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% PYTHIA 6.4 physics description and manual
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\begin{document}
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\begin{flushright}
hep-ph/0603175\\
LU TP 06--13\\
FERMILAB-PUB-06-052-CD-T\\
March 2006
\end{flushright}
\vspace{\fill}
\begin{center}
\tbf{{\Huge P}{\LARGE YTHIA}~~{\Huge 6.4}}\\[5mm]
\tbf{{\LARGE Physics and Manual}} \\[10mm]
{\large\bf Torbj\"orn Sj\"ostrand}\\[1mm]
Department of Theoretical Physics, Lund University,\\
S\"olvegatan 14A, S-223 62 Lund, Sweden\\
E-mail: \texttt{torbjorn@thep.lu.se}\\[4mm]
{\large\bf Stephen Mrenna}\\[1mm]
Computing Division, Simulations Group,
Fermi National Accelerator Laboratory,\\
MS 234, Batavia, IL 60510, USA\\
E-mail: \texttt{mrenna@fnal.gov}\\[4mm]
{\large\bf Peter Skands}\\[1mm]
Theoretical Physics Department,
Fermi National Accelerator Laboratory,\\
MS 106, Batavia, IL 60510, USA\\
E-mail: \texttt{skands@fnal.gov}
\end{center}
\vspace{\fill}
\begin{center}
{\bf Abstract}\\[2ex]
\begin{minipage}{\abstwidth}
The \textsc{Pythia} program can be used to generate high-energy-physics
`events', i.e.\ sets of outgoing particles produced in the interactions
between two incoming particles. The objective is to provide as accurate
as possible a representation of event properties in a wide range of
reactions, within and beyond the Standard Model, with emphasis on those
where strong interactions play a r\^ole, directly or indirectly, and
therefore multihadronic final states are produced. The physics is then
not understood well enough to give an exact description; instead the
program has to be based on a combination of analytical results and
various QCD-based models. This physics input is summarized here, for
areas such as hard subprocesses, initial- and final-state parton showers,
underlying events and beam remnants, fragmentation and decays, and much
more. Furthermore, extensive information is provided on all program
elements: subroutines and functions, switches and parameters, and particle
and process data. This should allow the user to tailor the generation task
to the topics of interest.\\
The code and further information may be found on the \textsc{Pythia}
web page:\\
\texttt{http://www.thep.lu.se/}$\sim$\texttt{torbjorn/Pythia.html}.
\end{minipage}
\end{center}
\vspace{\fill}
\phantom{dummy}
\clearpage
\phantom{dummy}
\vspace{\fill}
\begin{minipage}{14cm}
\begin{flushright}
\textit{\LARGE Dedicated to}\\[5mm]
\textit{\LARGE the Memory of}\\[9mm]
\textbf{\textit{\Huge Bo Andersson}}\\[10mm]
\textit{\Large 1937 -- 2002}\\[5mm]
\textit{\Large originator, inspirator}
\end{flushright}
\end{minipage}
\vspace{\fill}
\vspace{\fill}
\phantom{dummy}
\clearpage
\section*{Preface}
%\addcontentsline{toc}{section}{Preface}
The {\Py} program is frequently used for event generation
in high-energy physics. The emphasis is on multiparticle production
in collisions between elementary particles. This in particular means
hard interactions in $\ee$, $\pp$ and $\ep$ colliders, although
also other applications are envisaged. The program is intended to
generate complete events, in as much detail as
experimentally observable ones, within the bounds of our current
understanding of the underlying physics. Many of the components of
the program represent original research, in the sense that models
have been developed and implemented for a number of aspects not
covered by standard theory.
Event generators often have a reputation for being `black boxes';
if nothing else, this report should provide you with a glimpse of
what goes on inside the program. Some such understanding may be of
special interest for new users, who have no background in the field.
An attempt has been made to structure the report sufficiently well
so that many of the sections can be read independently of each other,
so you can pick the sections that interest you. We have tried to
keep together the physics and the manual sections on specific
topics, where practicable.
A large number of persons should be thanked for their contributions.
Bo Andersson and G\"osta Gustafson are the originators of the Lund
model, and strongly influenced the early development of related
programs. (Begun with {\Je} in 1978, now fused with {\Py}.)
Hans-Uno Bengtsson is the originator of the {\Py} program. Mats
Bengtsson is the main author of the old final-state parton-shower
algorithm. Patrik Ed\'en has contributed an improved popcorn scenario
for baryon production. Maria van Zijl has helped develop the original
multiple-interactions scenarios, Christer Friberg the expanded
photon physics machinery, Emanuel Norrbin the new matrix-element matching
of the final-state parton shower algorithm and the handling of low-mass
strings, Leif L\"onnblad the Bose--Einstein models, and Gabriela Miu the
matching of initial-state showers. Stefan Wolf provided an implementation
of onium production in NRQCD.
Further bug reports, smaller pieces of code and general comments on the
program have been obtained from users too numerous to be mentioned here,
but who are all gratefully acknowledged. To write programs of this size
and complexity would be impossible without a strong support and user
feedback. So, if you find errors, please let us know.
The moral responsibility for any remaining errors clearly rests with
the authors. However, kindly note that this is a `University World'
product, distributed `as is', free of charge, without any binding
guarantees. And always remember that the program does not represent a
dead collection of established truths, but rather one of many possible
approaches to the problem of multiparticle production in high-energy
physics, at the frontline of current research. Be critical!
\clearpage
\tableofcontents
\clearpage
\pagestyle{plain}
\setcounter{page}{1}
\section{Introduction}
Multiparticle production is the most characteristic feature
of current high-energy physics. Today, observed particle
multiplicities are typically between ten and a hundred, and
with future machines this range will be extended upward.
The bulk of the multiplicity is found in jets, i.e.\ in
collimated bunches of hadrons (or decay products of hadrons) produced
by the hadronization of partons, i.e.\ quarks and gluons. (For some
applications it will be convenient to extend the parton concept
also to some non-coloured but showering particles, such as
electrons and photons.)
\subsection{The Complexity of High-Energy Processes}
To first approximation, all processes have a simple structure at
the level of interactions between the fundamental objects of nature,
i.e.\ quarks, leptons and gauge bosons. For instance, a lot can
be understood about the structure of hadronic events at LEP just
from the `skeleton' process $\ee \to \Z^0 \to \q \qbar$.
Corrections to this picture can be subdivided, arbitrarily but
conveniently, into three main classes.
Firstly, there are bremsstrahlung-type modifications, i.e.\ the
emission of additional final-state particles by branchings such as
$\e \to \e \gamma$ or $\q \to \q \g$. Because of the largeness of
the strong coupling constant $\alphas$, and because of the presence
of the triple gluon vertex, QCD emission off quarks and gluons is
especially prolific. We therefore speak about `parton showers',
wherein a single initial parton may give rise to a whole bunch of
partons in the final state. Also photon emission may give sizable
effects in $\ee$ and $\ep$ processes. The bulk of the bremsstrahlung
corrections are universal, i.e.\ do not depend on the details of
the process studied, but only on one or a few key numbers, such as
the momentum transfer scale of the process. Such universal
corrections may be included to arbitrarily high orders, using a
probabilistic language. Alternatively, exact calculations of
bremsstrahlung corrections may be carried out order by order in
perturbation theory, but rapidly the calculations then become
prohibitively complicated and the answers correspondingly
lengthy.
Secondly, we have `true' higher-order corrections, which involve a
combination of loop graphs and the soft parts of the
bremsstrahlung graphs above, a combination needed to
cancel some divergences. In a complete description it is
therefore not possible to consider bremsstrahlung separately,
as assumed here. The
necessary perturbative calculations are usually very difficult;
only rarely have results been presented that include more than one
non-`trivial' order, i.e.\ more than one loop.
As above, answers
are usually very lengthy, but some results are sufficiently simple
to be generally known and used, such as the running of $\alphas$, or
the correction factor $1 + \alphas/\pi + \cdots$ in the partial
widths of $\Z^0 \to \q \qbar$ decay channels. For high-precision
studies it is imperative to take into account the results of
loop calculations, but usually effects are minor for the qualitative
aspects of high-energy processes.
Thirdly, quarks and gluons are confined. In the two points above,
we have used a perturbative language to describe the short-distance
interactions of quarks, leptons and gauge bosons. For leptons
and colourless bosons this language is sufficient. However, for
quarks and gluons it must be complemented with the structure of
incoming hadrons, and a picture for the hadronization process,
wherein the coloured partons are transformed into jets of colourless
hadrons, photons and leptons. The hadronization can be further
subdivided into fragmentation and decays, where the former describes
the way the creation of new quark-antiquark pairs can break up a
high-mass system into lower-mass ones, ultimately hadrons. (The word
`fragmentation' is also sometimes used in a broader sense, but we
will here use it with this specific meaning.) This process is still
not yet understood from first principles, but has to be based on models.
In one sense, hadronization effects are overwhelmingly large, since
this is where the bulk of the multiplicity comes from. In another
sense, the overall energy flow of a high-energy event is mainly
determined by the perturbative processes, with only a minor additional
smearing caused by the hadronization step. One may therefore pick
different levels of ambition, but in general detailed studies require
a detailed modelling of the hadronization process.
The simple structure that we started out with has now become
considerably more complex --- instead of maybe two final-state
partons we have a hundred final particles. The original physics
is not gone, but the skeleton process has been dressed up and
is no longer directly visible. A direct comparison between theory
and experiment is therefore complicated at best, and
impossible at worst.
\subsection{Event Generators}
It is here that event generators come to the rescue. In an event
generator, the objective striven for is to use computers to generate
events as detailed as could be observed by a perfect detector.
This is not done in one step, but rather by `factorizing' the full
problem into a number of components, each of which can be handled
reasonably accurately. Basically, this means that the hard process
is used as input to generate bremsstrahlung corrections, and that
the result of this exercise is thereafter left to hadronize. This
sounds a bit easier than it really is --- else this report would
be a lot thinner. However, the basic idea is there: if the
full problem is too complicated to be solved in one go, it may be
possible to
subdivide it into smaller tasks of more manageable proportions.
In the actual generation procedure, most steps therefore involve
the branching of one object into two, or at least into a very small
number, with the daughters free to branch in their turn. A lot of
book-keeping is involved, but much is of a repetitive nature, and
can therefore be left for the computer to handle.
As the name indicates, the output of an event generator should be
in the form of `events', with the same average behaviour and the
same fluctuations as real data. In the data, fluctuations arise from
the quantum mechanics of the underlying theory. In
generators, Monte Carlo techniques are used to select all relevant
variables according to the desired probability distributions,
and thereby ensure (quasi-)randomness in the final events.
Clearly some loss of information is entailed: quantum mechanics is
based on amplitudes, not probabilities. However, only very rarely
do (known) interference phenomena appear that cannot be cast in a
probabilistic language. This is therefore not a more restraining
approximation than many others.
Once there, an event generator can be used in many different ways.
The five main applications are probably the following:
\begin{Itemize}
\item To give physicists a feeling for the kind
of events one may expect/hope to find, and at what rates.
\item As a help in the planning of a new detector, so that detector
performance is optimized, within other constraints, for the
study of interesting physics scenarios.
\item As a tool for devising the analysis strategies that should
be used on real data, so that signal-to-background conditions are
optimized.
\item As a method for estimating detector acceptance corrections
that have to be applied to raw data, in order to extract the
`true' physics signal.
\item As a convenient framework within which to interpret the
observed phenomena in terms of a more fundamental
underlying theory (usually the Standard Model).
\end{Itemize}
Where does a generator fit into the overall analysis chain of an
experiment? In `real life', the machine produces interactions.
These events are observed by detectors, and the interesting ones
are written to tape by the
data acquisition system. Afterward the events may be reconstructed,
i.e.\ the electronics signals (from wire chambers, calorimeters, and
all the rest) may be
translated into a deduced setup of charged tracks or
neutral energy depositions, in the best of worlds with full knowledge
of momenta and particle species. Based on this cleaned-up
information, one may proceed with the physics analysis.
In the Monte Carlo world, the r\^ole of the machine, namely to produce
events, is taken by the event generators described in this report.
The behaviour of the detectors --- how particles produced by the
event generator traverse the detector, spiral in magnetic
fields, shower in calorimeters, or sneak out through cracks, etc.\ ---
is simulated in programs such as \tsc{Geant} \cite{Bru89}.
Be warned that this latter activity is sometimes called event simulation,
which is somewhat unfortunate
since the same words could equally well be applied to what, here, we
call event generation. A more appropriate term is detector
simulation. Ideally, the output of this simulation has exactly the
same format as the real data recorded by the detector, and can
therefore be put through the same event reconstruction and physics
analysis chain, except that here we know what the `right answer'
should be, and so can see how well we are doing.
Since the full chain of detector simulation and event
reconstruction is very
time-consuming, one often does `quick and dirty' studies in
which these steps are skipped entirely, or at least replaced by
very simplified procedures which only take into account the geometric
acceptance of the detector and other trivial effects. One may then
use the output of the event generator directly in the physics studies.
There are still many holes in our understanding of the full event
structure, despite an impressive amount of work and detailed
calculations. To put together a generator therefore involves making
a choice on what to include, and how to include it. At best, the
spread between generators can be used to give some impression of
the uncertainties involved. A multitude of approximations will
be discussed in the main part of this report, but already here
is should be noted that many major approximations are related to
the almost complete neglect of non-`trivial' higher-order effects,
as already mentioned. It can therefore only be hoped that
the `trivial' higher order parts give the bulk of the experimental
behaviour. By and large, this seems to be the case; for $\ee$
annihilation it even turns out to be a very good approximation.
The necessity to make compromises has one major implication:
to write a good event generator is an art, not an exact science.
It is therefore essential not to blindly trust the
results of any single event generator, but always to make several
cross-checks. In addition, with computer programs of tens of
thousands of lines, the question is not whether bugs exist, but how
many there are, and how critical their positions.
Further, an event generator cannot be thought of as all-powerful,
or able to give intelligent answers to ill-posed questions;
sound judgement and some understanding of a
generator are necessary prerequisites for successful use. In spite
of these limitations, the event-generator approach is the most
powerful tool at our disposal if we wish to gain a detailed and
realistic understanding of physics at current or future high-energy
colliders.
\subsection{The Origins of the Current Program}
Over the years, many event generators have appeared. A recent
comprehensive overview is the Les Houches
guidebook to Monte Carlo event generators \cite{Dob04}. Surveys of
generators for $\ee$ physics in general and LEP in particular
may be found in \cite{Kle89,Sjo89,Kno96,Lon96,Bam00}, for high-energy
hadron--hadron ($\pp$) physics in \cite{Ans90,Sjo92,Kno93,LHC00},
and for $\ep$ physics in \cite{HER92,HER99}. We refer the reader
to those for additional details and references. In this particular
report, the two closely connected programs {\Je} and {\Py}, now
merged under the {\Py} label, will be described.
{\Je} has its roots in the efforts of the Lund
group to understand the hadronization process, starting in the late
seventies \cite{And83}. The so-called string fragmentation model
was developed as an explicit and detailed framework, within which
the long-range confinement forces are allowed to distribute the
energies and flavours of a parton configuration among a collection
of primary hadrons, which subsequently may decay further. This model,
known as the Lund string model, or `Lund' for short, contained a
number of specific predictions, which were confirmed by data from
$\e^+\e^-$ annihilation around 30 \GeV\ at PETRA and PEP,
whence the model gained a widespread acceptance.
The Lund string model is still today the most elaborate
and widely used fragmentation model at our disposal. It remains
at the heart of the {\Py} program.
In order to predict the shape of events at PETRA/PEP, and to
study the fragmentation process in detail, it was necessary to start
out from the partonic configurations that were to fragment.
The generation of complete $\ee$ hadronic events was therefore
added, originally based on simple $\gamma$ exchange and
first-order QCD matrix elements, later extended to full $\gammaZ$
exchange with first-order initial-state QED radiation and
second-order QCD matrix elements. A number of utility routines
were also provided early on, for everything from event listing
to jet finding.
By the mid-eighties it was clear that the pure matrix-element approach
had reached the limit of its usefulness, in the sense that it could not
fully describe the exclusive multijet topologies of the data.
(It is still useful for inclusive descriptions, like the optimized
perturbation theory discussed in section \ref{sss:optimizedpt},
and in combination with renormalon contributions \cite{Dok97}.)
Therefore a parton-shower description was developed \cite{Ben87a} as
an alternative to the higher-order matrix-element one. (Or rather
as a complement, since the trend over the years has been towards
the development of methods to marry the two approaches.)
The combination of parton showers and string fragmentation has been
very successful, and has formed the main approach to the description
of hadronic $\Z^0$ events.
This way, the {\Je} code came to cover the four main areas of
fragmentation, final-state parton showers, $\ee$ event generation
and general utilities.
The successes of string fragmentation in $\ee$ made it interesting
to try to extend this framework to other processes, and explore
possible physics consequences. Therefore a number of
other programs were written, which combined a process-specific
description of the hard interactions with the general fragmentation
framework of {\Je}. The {\Py} program
evolved out of early studies on fixed-target proton--proton
processes, addressed mainly at issues related to string drawing.
With time, the interest shifted towards hadron collisions at higher
energies, first to the SPS $\ppbar$ collider, and later to the
Tevatron, SSC and LHC, in the context of a number of workshops in the
USA and Europe. Parton showers were added, for final-state radiation
by making use of the {\Je} routine, for initial-state one by the
development of the concept of `backwards evolution', specifically for
{\Py} \cite{Sjo85}. Also a framework was developed for minimum-bias
and underlying events \cite{Sjo87a}.
Another main change was the introduction of an increasing
number of hard processes, within the Standard Model and beyond.
A special emphasis was put on the search for the Standard Model
Higgs, in different mass ranges and in different channels, with due
respect to possible background processes.
The bulk of the machinery developed for hard processes actually
depended little on the choice of initial state, as long as the
appropriate parton distributions were there for the incoming
partons and particles. It therefore made sense to extend the
program from being only a $\pp$ generator to working also for
$\ee$ and $\ep$. This process was completed in 1991,
again spurred on by physics workshop activities. Currently
{\Py} should therefore work well for a selection
of different possible incoming beam particles.
An effort independent of the Lund group activities got going to
include supersymmetric event simulation in {\Py}. This resulted
in the \tsc{SPythia} program \cite{Mre97}.
While {\Je} was independent of {\Py} until 1996, their ties
had grown much stronger over the years, and the border-line
between the two programs had become more and more artificial.
It was therefore decided to merge the two, and also include the
\tsc{SPythia} extensions, starting from {\Py}~6.1. The different
origins in part still are reflected in this manual,
%but the strive is towards a seamless merger.
%referee: misprint
but the striving is towards a seamless merger.
Among the most recent developments, primarily intended for Tevatron and LHC
physics studies, is the introduction of `interleaved evolution' in
\Py\ 6.3, with new $\pT$-ordered parton showers
and a more sophisticated framework for minimum-bias and underlying events
\cite{Sjo04,Sjo04a}. The possibilities for studying physics beyond
the Standard Model have also been extended significantly, to include
supersymmetric models with $R$-parity violation, Technicolor models,
$\Z'/\W'$ models, as well as models with (Randall--Sundrum) extra dimensions.
This still only includes the models available internally in \Py.
Versatility is further enhanced by the addition of an interface to external
user processes, according to the Les Houches Accord (LHA) standard
\cite{Boo01}, and by interfaces to SUSY RGE and decay packages via the
SUSY Les Houches Accord (SLHA) \cite{Ska03}.
The tasks of including new processes, and of improving the simulation
of parton showers and other aspects of already present processes, are
never-ending. Work therefore continues apace.
\subsection{About this Report}
As we see, {\Je} and {\Py} started out as very
ideologically motivated programs, developed to study specific
physics questions in enough detail that explicit predictions
could be made for experimental quantities. As it was recognized
that experimental imperfections could distort the basic predictions,
the programs were made available for general use by experimentalists.
It thus became feasible to explore the models in more detail
than would otherwise have been possible. As time went by, the
emphasis came to shift somewhat, away from the original strong
coupling to a specific fragmentation model, towards a description of
high-energy multiparticle production processes in general.
Correspondingly, the use expanded from being one of just comparing
data with specific model predictions, to one of extensive use for
the understanding of detector performance, for the derivation of
acceptance correction factors, for the prediction of physics at
future high-energy accelerators, and for the design of related
detectors.
While the ideology may be less apparent, it is still there, however.
This is not something unique to the programs discussed here,
but inherent in any event generator, or at least any generator that
attempts to go beyond the simple parton level skeleton description
of a hard process. Do not accept the myth that everything available
in Monte Carlo form represents ages-old common knowledge, tested
and true. Ideology is present by commissions or omissions
%in any number of details. A programs like {\Py} represents
%typo
in any number of details. A program like {\Py} represents
a major amount of original physics research, often on complicated
topics where no simple answers are available.
As a (potential) program user you must be
aware of this, so that you can form your own opinion, not just about
what to trust and what not to trust, but also how much to trust a given
prediction, i.e.\ how uncertain it is likely to be.
{\Py} is particularly well endowed in this respect, since a
number of publications exist where most of the relevant physics is
explained in considerable detail. In fact, the problem may rather be
the opposite, to find the relevant information among all the possible
places. One main objective of the current report is therefore to
collect much of this information in one single place. Not all the
material found in specialized papers is reproduced, by a wide margin,
but at least enough should be found here to understand the general
picture and to know where to go for details.
The official reference for \Py\ is therefore the current report. It
is intended to update and extend the previous round of published
physics descriptions and program manuals \cite{Sjo01,Sjo01a,Sjo03a}.
Further specification could include a statement of
the type `We use {\Py} version X.xxx'. (If you are a {\LaTeX} fan,
you may want to know that the program name in this report has been
generated by the command \verb+\textsc{Pythia}+.)
Kindly do not refer to {\Py} as `unpublished', `private communication'
or `in preparation': such phrases are incorrect and only create
unnecessary confusion.
In addition, remember that many of the individual physics components
are documented in separate publications. If some of these contain
ideas that are useful to you, there is every reason to cite them.
A reasonable selection would vary as a function of the physics you are
studying. The criterion for which to pick should be simple: imagine
that a Monte Carlo implementation had not been available. Would you
then have cited a given paper on the grounds of its physics contents
alone? If so, do not punish the extra effort of turning these ideas
into publicly available software. (Monte Carlo manuals are good for
nothing in the eyes of many theorists, so often only the acceptance
of `mainstream' publications counts.) Here follows a list of some
main areas where the $\Py$ programs contain original research:
\paragraph{Fragmentation/Hadronization:}
\begin{Itemize}
\item The string fragmentation model \cite{And83,And98}.
\item The string effect \cite{And80}.
\item Baryon production (diquark/popcorn) \cite{And82,And85,Ede97}.
\item Fragmentation of systems with string junctions \cite{Sjo03}.
\item Small-mass string fragmentation \cite{Nor98}.
\item Fragmentation of multiparton systems \cite{Sjo84}.
\item Colour rearrangement \cite{Sjo94a} and Bose-Einstein effects
\cite{Lon95}.
\item Fragmentation effects on $\alphas$ determinations \cite{Sjo84a}.
\end{Itemize}
\paragraph{Parton Showers:}
\begin{Itemize}
\item Initial-state parton showers ($Q^2$-ordering) \cite{Sjo85,Miu99}.
\item Final-state parton showers ($Q^2$-ordering) \cite{Ben87a,Nor01}.
\item Initial-state parton showers ($\pTs$-ordering) \cite{Sjo04a}.
\item Final-state parton showers ($\pTs$-ordering) \cite{Sjo04a}.
\item Photon radiation from quarks \cite{Sjo92c}
\end{Itemize}
\paragraph{DIS and photon physics:}
\begin{Itemize}
\item Deeply Inelastic Scattering \cite{And81a,Ben88}.
\item Photoproduction \cite{Sch93a}, $\gamma\gamma$ \cite{Sch94a}
and $\gast\p/\gast\gamma/\gast\gast$ \cite{Fri00} physics.
\item Parton distributions of the photon \cite{Sch95,Sch96}.
\end{Itemize}
\paragraph{Beyond the Standard Model physics:}
\begin{Itemize}
\item Supersymmetry \cite{Amb96,Mre99a},
with $R$-parity violation \cite{Ska01,Sjo03}.
\item Technicolor \cite{Lan02a}.
\item Extra dimensions \cite{Bij01}
\item $\Z'$ models \cite{Lyn00}
\end{Itemize}
\paragraph{Other topics:}
\begin{Itemize}
\item Colour flow in hard scatterings \cite{Ben84}.
\item Elastic and diffractive cross sections \cite{Sch94}.
\item Minijets, underlying event, and minimum-bias
(multiple parton--parton interactions)
\cite{Sjo87a,Sjo04,Sjo04a}.
\item Rapidity gaps \cite{Dok92}.
\item Jet clustering in $k_{\perp}$ \cite{Sjo83}.
\end{Itemize}
In addition to a physics survey, the current report also contains a
complete manual for the program. Such manuals have always been
updated and distributed jointly with the programs, but have grown in
size with time. A word of warning may therefore be in place.
The program description is fairly
lengthy, and certainly could not be absorbed in one sitting. This is
not even necessary, since all switches and parameters are provided
with sensible default values, based on our best understanding (of
the physics, and of what you expect to happen if you do not
specify any options). As a new user, you can therefore disregard all
the fancy options, and just run the program with a minimum ado.
Later on, as you gain experience, the options that seem useful can
be tried out. No single user is ever likely to find need for more than
a fraction of the total number of possibilities available,
yet many of them have been added to meet specific user requests.
In some instances, not even this report will provide you with all the
information you desire. You may wish to find out about recent versions
of the program, know about related software, pick up a few sample
main programs to get going, or get hold of related physics papers.
Some such material can be found on the {\Py} web page:\\
\texttt{http://www.thep.lu.se/}$\sim$\texttt{torbjorn/Pythia.html}.
\subsection{Disclaimer}
At all times it should be remembered that this is not a commercial
product, developed and supported by professionals. Instead it is
a `University World' product, developed by a very few physicists
(mainly the current first author) originally for their own needs,
and supplied to other physicists on an `as-is' basis, free of charge.
(It is protected by copyright, however, so is not `free software'
in the nowadays common meaning. This is not intended to stifle
research, but to make people respect some common-sense
`intellectual property' rights: code should not be `borrowed'
and redistributed in such a form that credit would not go to the
people who did the work.)
No guarantees are
therefore given for the proper functioning of the program, nor for
the validity of physics results. In the end, it is always up to you
to decide for yourself whether to trust a given result or not. Usually
this requires comparison either with analytical results or with
results of other programs, or with both. Even this is not necessarily
foolproof: for instance, if an error
is made in the calculation of a matrix element for a given process,
this error will be propagated both into the analytical results based
on the original calculation and into all the event generators which
subsequently make use of the published formulae. In the end, there
is no substitute for a sound physics judgement.
This does not mean that you are all on your own, with a program
nobody feels responsible for. Attempts are made to check processes as
carefully as possible, to write programs that do not invite
unnecessary errors, and to provide a detailed and accurate
documentation. All of this while maintaining the full power and
flexibility, of course, since the physics must always take precedence
in any conflict of interests. If nevertheless any errors or
unclear statements are found, please do communicate them to
one of the authors. Every attempt will be made to solve problems
as soon as is reasonably possible, given that this support is
by a few persons, who mainly have other responsibilities.
However, in order to make debugging at all possible, we request
that any sample code you want to submit as evidence be completely
self-contained, and peeled off from all irrelevant aspects. Use
simple write statements or the {\Py} histogramming routines to make
your point. Chances are that, if the error cannot be reproduced
by fifty lines of code, in a main program linked only to {\Py},
the problem is sitting elsewhere. Numerous errors have been caused
by linking to other (flawed) libraries, e.g.\ collaboration-specific
frameworks for running {\Py}. Then you should put the blame elsewhere.
\subsection{Appendix: The Historical Pythia}
The `{\Py}' label may need some explanation.
The myth tells how Apollon, the God of Wisdom, killed the powerful
dragon-like monster Python, close to the village of Delphi in Greece.
To commemorate this victory, Apollon founded the Pythic Oracle in
Delphi, on the slopes of Mount Parnassos. Here men could come to
learn the will of the Gods and the course of the future. The oracle
plays an important r\^ole in many of the other Greek myths, such
as those of Heracles and of King Oedipus.
Questions were to be put to the Pythia, the `Priestess' or
`Prophetess' of the Oracle. In fact, she was a local woman,
usually a young maiden, of no particular religious schooling.
Seated on a tripod, she inhaled the obnoxious vapours that
seeped up through a crevice in the ground. This brought her
to a trance-like state, in which she would scream seemingly
random words and sounds. It was the task of the professional
priests in Delphi to record those utterings and edit them into
the official Oracle prophecies, which often took the form of
poems in perfect hexameter. In fact, even these edited replies
were often less than easy to interpret. The Pythic oracle
acquired a reputation for ambiguous answers.
The Oracle existed already at the beginning of the historical
era in Greece, and was universally recognized as the foremost
religious seat. Individuals and city states came to consult, on
everything from cures for childlessness to matters of war. Lavish
gifts allowed the temple area to be built and decorated. Many
states supplied their own treasury halls, where especially beautiful
gifts were on display. Sideshows included the Omphalos,
a stone reputedly marking the centre of the Earth, and the Pythic
games, second only to the Olympic ones in importance.
Strife inside Greece eventually led to a decline in the power of
the Oracle. A serious blow was dealt when the Oracle of Zeus Ammon
(see below) declared Alexander the Great
to be the son of Zeus. The Pythic Oracle lived on, however, and was
only closed by a Roman Imperial decree in 390 \tsc{ad}, at a time
when Christianity was ruthlessly destroying any religious opposition.
Pythia then had been at the service of man and Gods for a
millennium and a half.
The r\^ole of the Pythic Oracle prophecies on the course of history
is nowhere better described than in `The Histories' by Herodotus
\cite{HerBC}, the classical and captivating description of the
Ancient World at the time of the Great War between Greeks and
Persians. Especially famous is the episode with King Croisus
of Lydia. Contemplating a war against the upstart Persian
Empire, he resolves to ask an oracle what the outcome of a potential
battle would be. However, to have some guarantee for the
veracity of any prophecy, he decides to send embassies to all the
renowned oracles of the known World. The messengers are instructed
to inquire the various divinities, on the hundredth day after
their departure, what King Croisus is doing at that very moment.
{}From the Pythia the messengers bring back the reply
\begin{em}\begin{verse}
I know the number of grains of sand as well as the expanse of
the sea, \\
And I comprehend the dumb and hear him who does not speak, \\
There came to my mind the smell of the hard-shelled turtle, \\
Boiled in copper together with the lamb, \\
With copper below and copper above.
\end{verse}\end{em}
The veracity of the Pythia is thus established by the crafty ruler,
who had waited until the appointed day, slaughtered a turtle and a
lamb, and boiled them together in a copper cauldron with a
copper lid. Also the Oracle of Zeus Ammon in the Libyan desert
is able to give a correct reply (lost to posterity), while all
others fail. King Croisus now sends a second embassy to Delphi,
inquiring after the outcome of a battle against the Persians.
The Pythia answers
\begin{em}\begin{verse}
If Croisus passes over the Halys he will dissolve a great Empire.
\end{verse}\end{em}
Taking this to mean he would win, the King collects his army and
crosses the border river, only to suffer a crushing defeat and
see his Kingdom conquered. When the victorious King Cyrus allows
Croisus to send an embassy to upbraid the Oracle, the God Apollon
answers through his Prophetess that he has correctly predicted the
destruction of a great empire --- Croisus' own --- and that he
cannot be held responsible if people choose to interpret the
Oracle answers to their own liking.
The history of the {\Py} program is neither as long nor as
dignified as that of its eponym. However, some points of contact
exist. You must be very careful when you formulate the questions:
any ambiguities will corrupt the reply you get. And you must be even
more careful not to misinterpret the answers; in particular not to pick
the interpretation that suits you before considering the alternatives.
Finally, even a perfect God has servants that are only human: a priest
might mishear the screams of the Pythia and therefore produce an
erroneous oracle reply; the current authors might unwittingly let a
bug free in the program {\Py}.
\clearpage
\section{Physics Overview}
In this section we will try to give an overview of the main physics
features of {\Py}, and also to introduce some
terminology. The details will be discussed in subsequent sections.
For the description of a typical high-energy event, an event
generator should contain a simulation of several physics aspects.
If we try to follow the evolution of an event in some semblance of
a time order, one may arrange these aspects as follows:
\begin{Enumerate}
\item Initially two beam particles are coming in towards each other.
Normally each particle is characterized by a set of parton
distributions, which defines the partonic substructure in terms
of flavour composition and energy sharing.
\item One shower initiator parton from each beam starts off
a sequence of branchings, such as $\q \to \q \g$, which build up
an initial-state shower.
\item One incoming parton from each of the two showers
enters the hard process, where then a number of
outgoing partons are produced, usually two.
It is the nature of this process that determines the main
characteristics of the event.
\item The hard process may produce a set of short-lived resonances,
like the $\Z^0/\W^{\pm}$ gauge bosons, whose decay to normal
partons has to be considered in close association with the
hard process itself.
\item The outgoing partons may branch, just like the incoming did,
to build up final-state showers.
\item In addition to the hard process considered above, further
semihard interactions may occur between the other partons
of two incoming hadrons.
\item When a shower initiator is taken out of a beam particle,
a beam remnant is left behind. This remnant may have
an internal structure, and a net colour charge that relates
it to the rest of the final state.
\item The QCD confinement mechanism ensures that the outgoing quarks
and gluons are not observable, but instead fragment to colour
neutral hadrons.
\item Normally the fragmentation mechanism can be seen as occurring
in a set of separate colour singlet subsystems, but
interconnection effects such as colour rearrangement or
Bose--Einstein may complicate the picture.
\item Many of the produced hadrons are unstable and decay further.
\end{Enumerate}
Conventionally, only quarks and gluons are counted as partons, while
leptons and photons are not. If pushed {\it ad absurdum} this may
lead to
some unwieldy terminology. We will therefore, where it does not matter,
speak of an electron or a photon in the `partonic' substructure of an
electron, lump branchings $\e \to \e \gamma$ together with other
`parton shower' branchings such as $\q \to \q \g$, and so on. With
this notation, the division into the above ten points applies equally
well to an interaction between two leptons, between a lepton and a
hadron, and between two hadrons.
In the following sections, we will survey the above ten aspects,
not in the same order as given here, but rather in the order in
which they appear in the program execution, i.e.\ starting with the
hard process.
\subsection{Hard Processes and Parton Distributions}
In the original {\Je} code, only two hard processes were available. The
first and main one is $\ee \to \gammaZ \to \q \qbar$. Here the `$*$' of
$\gamma^*$ is used to denote that the photon must be off the mass shell.
The distinction is of some importance, since a photon on the mass shell
cannot decay. Of course also the $\Z^0$ can be off the mass shell,
but here the distinction is less relevant (strictly speaking,
a $\Z^0$ is always off the mass shell). In the following we may not
always use `$*$' consistently, but the rule of thumb is to use a `$*$'
only when a process is not kinematically possible for a particle of
nominal mass. The quark $\q$ in the final state of
$\ee \to \gammaZ \to \q \qbar$
may be $\u$, $\d$, $\s$, $\c$, $\b$ or $\t$; the flavour in
each event is picked at random, according to the relative couplings,
evaluated at the hadronic c.m.\ energy. Also the angular distribution of
the final $\q \qbar$ pair is included. No parton-distribution functions
are needed.
The other original {\Je} process is a routine to generate $\g \g \g$ and
$\gamma \g \g$ final states, as expected in onium 1$^{--}$ decays
such as $\Upsilon$. Given the large top mass, toponium decays weakly much
too fast for these processes to be of any interest, so therefore no new
applications are expected.
\subsubsection{Hard Processes}
The current {\Py} contains a much richer selection, with around 300
different hard processes. These may be classified in
many different ways.
One is according to the number of final-state objects: we speak of
`$2 \to 1$' processes, `$2 \to 2$' ones, `$2 \to 3$' ones, etc.
This aspect is very relevant from a programming point of view:
the more particles in the final state, the more complicated the
phase space and therefore the whole generation procedure. In fact,
{\Py} is optimized for $2 \to 1$ and $2 \to 2$ processes.
There is currently no generic treatment of processes with three or
more particles in the final state, but rather a few different
machineries, each tailored to the pole structure of a specific class of
graphs.
Another classification is according to the physics scenario. This will
be the main theme of section \ref{s:pytproc}. The following major
groups may be distinguished:
\begin{Itemize}
\item Hard QCD processes, e.g.\ $\q \g \to \q \g$.
\item Soft QCD processes, such as diffractive and elastic scattering,
and minimum-bias events. Hidden in this class is also process 96,
which is used internally for the merging of soft and hard physics,
and for the generation of multiple interactions.
\item Heavy-flavour production, both open and hidden, e.g.\
$\g \g \to \t \tbar$ and $\g \g \to \Jpsi \g$.
\item Prompt-photon production, e.g.\ $\q \g \to \q \gamma$.
\item Photon-induced processes, e.g.\ $\gamma \g \to \q \qbar$.
\item Deeply Inelastic Scattering, e.g.\ $\q \ell \to \q \ell$.
\item $\W / \Z$ production, such as the $\ee \to \gammaZ$ or
$\q \qbar \to \W^+ \W^-$.
\item Standard Model Higgs production, where the Higgs is reasonably
light and narrow, and can therefore still be considered as a resonance.
\item Gauge boson scattering processes, such as $\W \W \to \W \W$,
when the Standard Model Higgs is so heavy and broad that resonant and
non-resonant contributions have to be considered together.
\item Non-standard Higgs particle production, within the framework
of a two-Higgs-doublet scenario with three neutral ($\hrm^0$, $\H^0$
and $\A^0$) and two charged ($\H^{\pm}$) Higgs states. Normally
associated with SUSY (see below), but does not have to be.
\item Production of new gauge bosons, such as a $\Z'$, $\W'$ and $\R$
(a horizontal boson, coupling between generations).
\item Technicolor production, as an alternative scenario to the
standard picture of electroweak symmetry breaking by a fundamental Higgs.
\item Compositeness is a possibility not only in the Higgs sector,
but may also apply to fermions, e.g.\ giving $\d^*$ and $\u^*$ production.
At energies below the threshold for new particle production, contact
interactions may still modify the standard behaviour.
\item Left--right symmetric models give rise to doubly charged Higgs
states, in fact one set belonging to the left and one to the right
{\bf SU(2)} gauge group. Decays involve right-handed $\W$'s and neutrinos.
\item Leptoquark ($\L_{\Q}$) production is encountered in some
beyond-the-Standard-Model scenarios.
\item Supersymmetry (SUSY) is probably the favourite scenario for
physics beyond the Standard Model. A rich set of processes are allowed,
already if one obeys $R$-parity conservation, and even more so if one
does not. The main supersymmetric machinery and process selection is
inherited from \textsc{SPythia}~\cite{Mre97}, however with many
improvements in the event generation chain. Many different SUSY
scenarios have been proposed, and the program is flexible enough to
allow input from several of these, in addition to the ones provided
internally.
\item The possibility of extra dimensions at low energies has been
a topic of much study in recent years, but has still not settled down to
some standard scenarios. Its inclusion into {\Py} is also only in a very
first stage.
\end{Itemize}
This is by no means a survey of all interesting physics. Also, within
the scenarios studied, not all contributing graphs have always been
included, but only the more important and/or more interesting ones.
In many cases, various approximations are involved in the matrix
elements coded.
\subsubsection{Resonance Decays}
\label{sss:resdecintro}
As we noted above, the bulk of the processes above are of the
$2 \to 2$ kind, with very few leading to the production of more
than two final-state particles.
This may be seen as a major limitation, and indeed is so
at times. However, often one can come quite far with only one
or two particles in the final state, since showers will add the
required extra activity. The classification may also be misleading
at times, since an $s$-channel resonance is considered as a single
particle, even if it is assumed always to decay into two final-state
particles. Thus the process
$\ee \to \W^+ \W^- \to \q_1 \qbar'_1 \, \q_2 \qbar'_2$ is classified
as $2 \to 2$, although the decay treatment of the $\W$ pair includes
the full $2 \to 4$ matrix elements (in the doubly resonant
approximation, i.e.\ excluding interference with non-$\W\W$
four-fermion graphs).
Particles which admit this close connection between the hard process
and the subsequent evolution are collectively called resonances in
this manual. It includes all particles in mass above the
$\b$ quark system, such as $\t$, $\Z^0$, $\W^{\pm}$, $\hrm^0$,
supersymmetric particles, and many more. Typically their decays are
given by electroweak physics, or physics beyond the Standard Model.
What characterizes a ({\Py}) resonance is that partial widths
and branching ratios can be calculated dynamically, as a function of
the actual mass of a particle. Therefore not only do branching ratios
change between an $\hrm^0$ of nominal mass 100 GeV and one of 200 GeV,
but also for a Higgs of nominal mass 200 GeV, the branching ratios
would change between an actual mass of 190 GeV and 210 GeV, say.
This is particularly relevant for reasonably broad resonances, and
in threshold regions. For an approach like this to work, it is
clearly necessary to have perturbative expressions available for all
partial widths.
Decay chains can become quite lengthy, e.g.\ for supersymmetric processes,
but follow a straight perturbative pattern. If the simulation is
restricted to only some set of decays, the corresponding cross section
reduction can easily be calculated. (Except in some rare cases where a
nontrivial threshold behaviour could complicate matters.) It is
therefore standard in {\Py} to quote cross sections with such reductions
already included. Note that the branching ratios of a particle is affected
also by restrictions made in the secondary or subsequent decays.
For instance, the branching ratio of $\hrm^0 \to \W^+ \W^-$, relative to
$\hrm^0 \to \Z^0 \Z^0$ and other channels, is changed if the allowed $\W$
decays are restricted.
The decay products of resonances are typically quarks, leptons, or other
resonances, e.g.\ $\W \to \q \qbar'$ or $\hrm^0 \to \W^+ \W^-$. Ordinary
hadrons are not produced in these decays, but only in subsequent
hadronization steps. In decays to quarks, parton showers are
automatically added to give a more realistic multijet structure, and
one may also allow photon emission off leptons. If the decay products
in turn are resonances, further decays are necessary. Often
spin information is available in resonance decay matrix elements.
This means that the angular orientations in the two decays of a
$\W^+ \W^-$ pair are properly correlated. In other cases, the
information is not available, and then resonances decay isotropically.
Of course, the above `resonance' terminology is arbitrary. A $\rho$,
for instance, could also be called a resonance, but not in the above
sense. The width is not perturbatively calculable, it decays to hadrons
by strong interactions, and so on. From a practical point of view, the
main dividing line is that the values of --- or a change in --- branching
ratios cannot affect the cross section of a process. For instance, if
one wanted to consider the decay $\Z^0 \rightarrow \c \cbar$,
with a $\D$ meson producing a lepton, not only
would there then be the problem of different leptonic branching ratios
for different $\D$'s (which means that fragmentation and decay
treatments would no longer decouple), but also that of additional
$\c \cbar$ pair production in parton-shower evolution, at a rate
that is unknown beforehand. In practice, it is therefore next to
impossible to force $\D$ decay modes in a consistent manner.
\subsubsection{Parton Distributions}
The cross section for a process $ij \to k$ is given by
\begin{equation}
\sigma_{ij \to k} = \int \d x_1 \int \d x_2 \, f^1_i(x_1) \,
f^2_j(x_2) \, \hat{\sigma}_{ij \to k} ~.
\end{equation}
Here $\hat{\sigma}$ is the cross section for the hard partonic process,
as codified in the matrix elements for each specific process.
For processes with many particles in the final state
it would be replaced by an
integral over the allowed final-state phase space. The $f^a_i(x)$ are
the parton-distribution functions, which describe the probability to
find a parton $i$ inside beam particle $a$, with parton $i$ carrying a
fraction $x$ of the total $a$ momentum. Actually, parton distributions
also depend on some momentum scale $Q^2$ that characterizes the hard
process.
Parton distributions are most familiar for hadrons, such as the
proton, which are inherently composite objects, made up of quarks and
gluons. Since we do not understand QCD, a derivation from first
principles
of hadron parton distributions does not yet exist, although some
progress is being made in lattice QCD studies. It is therefore
necessary to rely on parameterizations, where experimental data are
used in conjunction with the evolution equations for the $Q^2$
dependence, to pin down the parton distributions. Several different
groups have therefore produced their own fits, based on slightly
different sets of data, and with some variation in the theoretical
assumptions.
Also for fundamental particles, such as the electron, is it convenient
to introduce parton distributions. The function $f^{\e}_{\e}(x)$ thus
parameterizes the probability that the electron that takes part in the
hard process retains a fraction $x$ of the original energy, the rest
being radiated (into photons) in the initial state. Of course, such
radiation could equally well be made part of the hard interaction,
but the parton-distribution approach usually is much more convenient.
If need be, a description with fundamental electrons is recovered for
the choice $f_{\e}^{\e}(x, Q^2) = \delta(x-1)$. Note that, contrary to
the proton case, electron parton distributions are calculable from first
principles, and reduce to the $\delta$ function above for $Q^2 \to 0$.
The electron may also contain photons, and the photon may in its turn
contain quarks and gluons. The internal structure of the
photon is a bit of a problem, since the photon contains a point-like
part, which is perturbatively calculable, and a resolved part (with
further subdivisions), which is not. Normally, the photon
parton distributions are therefore parameterized, just as the hadron
ones. Since the electron ultimately contains quarks and gluons,
hard QCD processes like $\q \g \to \q \g$ therefore not only appear in
$\pp$ collisions, but also in $\ep$ ones (`resolved photoproduction')
and in $\ee$ ones (`doubly resolved 2$\gamma$ events'). The parton
distribution function approach here makes it much easier to reuse one
and the same hard process in different contexts.
There is also another kind of possible generalization. The two
processes $\q \qbar \to \gammaZ$, studied in hadron colliders,
and $\ee \to \gammaZ$, studied in $\ee$ colliders, are really
special cases of a common process, $\f \fbar \to \gammaZ$,
where $\f$ denotes a fundamental fermion, i.e.\ a quark, lepton or
neutrino. The whole structure is therefore only coded once, and
then slightly different couplings and colour prefactors are used,
depending on the initial state considered. Usually the
interesting cross section is a sum over several different initial
states, e.g.\ $\u \ubar \to \gammaZ$ and
$\d \dbar \to \gammaZ$ in a hadron collider. This kind of
summation is always implicitly done, even when not explicitly
mentioned in the text.
A final comment on parton distributions is that, in general, the
composite structure of hadrons allow for
\emph{multiple parton--parton scatterings} to occur, in which case
correllated parton distributions should be used to describe the multi-parton
structure of the incoming beams. This will be discussed in section
\ref{ss:overview:brmi}.
\subsection{Initial- and Final-State Radiation}
In every process that contains coloured and/or charged objects
in the initial or final state, gluon and/or photon radiation
may give large corrections to the overall topology of events.
Starting from a basic $2 \to 2$ process, this kind of corrections
will generate $2 \to 3$, $2 \to 4$, and so on, final-state
topologies. As the available energies are increased, hard
emission of this kind is increasingly important, relative to
fragmentation, in determining the event structure.
Two traditional approaches exist to the modelling of perturbative
corrections. One is the matrix-element method, in
which Feynman diagrams are calculated, order by order. In principle,
this is the correct approach, which takes into account exact
kinematics, and the full interference and helicity structure. The
only problem is that calculations become increasingly difficult in
higher orders, in particular for the loop graphs.
Only in exceptional cases have therefore more than one loop been
calculated in full, and often we do not have any loop corrections
at all at our disposal. On the other hand,
we have indirect but strong evidence that, in fact, the emission
of multiple soft gluons plays a significant r\^ole in building up the
event structure, e.g.\ at LEP, and this sets a limit to the
applicability of matrix elements.
Since the phase space available for gluon emission increases with
the available energy, the matrix-element approach becomes less
relevant for the full structure of events at higher energies.
However, the perturbative expansion is better behaved at higher
energy scales, owing to the running of $\alphas$. As a consequence,
inclusive measurements, e.g.\ of the rate of well-separated jets,
should yield more reliable results at high energies.
The second possible approach is the parton-shower one. Here an
arbitrary number of branchings of one parton into two (or more)
may be combined, to yield a description of multijet events,
with no explicit upper limit on the number of partons involved.
This is possible since the full matrix-element expressions are
not used, but only approximations derived by simplifying the
kinematics, and the interference and helicity structure. Parton showers
are therefore expected to give a good description of the substructure of
jets, but in principle the shower approach has limited predictive power
for the rate of well-separated jets (i.e.\ the 2/3/4/5-jet composition).
In practice, shower programs may be matched to first-order matrix
elements to describe the hard-gluon emission region reasonably well,
in particular for the $\ee$ annihilation process.
Nevertheless, the shower description is not optimal for
absolute $\alphas$ determinations.
Thus the two approaches are complementary in many respects,
and both have found use. Because of its simplicity
and flexibility, the parton-shower option is often the
first choice, while the full higher-order matrix elements one
(i.e.\ including loops) is mainly used for $\alphas$ determinations,
angular distribution of jets, triple-gluon vertex studies, and other
specialized studies. With improved calculational techniques and
faster computers, Born-level calculations have been pushed to
higher orders, and have seen increasing use.
Obviously, the ultimate goal would be to have an approach where the
best aspects of the two worlds are harmoniously married. This is
currently a topic of quite some study, with several new approaches
having emerged over the last few years.
\subsubsection{Matrix elements}
Matrix elements are especially made use of in the older {\Je}-originated
implementation of the process $\ee \to \gammaZ \to \q \qbar$.
For initial-state QED radiation, a first-order (un-exponentiated)
description has been adopted. This means that events are subdivided
into two classes, those where a photon is radiated above some
minimum energy, and those without such a photon. In the latter
class, the soft and virtual corrections have been lumped together
to give a total event rate that is correct up to one loop. This
approach worked fine at PETRA/PEP energies, but does not do so well
for the $\Z^0$ line shape, i.e.\ in regions where the cross section
is rapidly varying and high precision is strived for.
For final-state QCD radiation, several options are available. The
default is the parton-shower one (see below), but some matrix-elements
options also exist. In the definition of 3- or
4-jet events, a cut is introduced whereby it is required that
any two partons have an invariant mass bigger than some fraction
of the c.m.\ energy. 3-jet events which do not fulfil this
requirement are lumped with the 2-jet ones. The first-order
matrix-element option, which only contains 3- and 2-jet events
therefore involves no ambiguities. In second order, where also
4-jets have to be considered, a main issue is what to do with
4-jet events that fail the cuts. Depending on the choice of
recombination scheme, whereby the two nearby partons are joined
into one, different 3-jet events are produced. Therefore the
second-order differential 3-jet rate has been the subject of
some controversy, and the program actually contains two
different implementations.
By contrast, the normal {\Py} event generation machinery does not
contain any full higher-order matrix elements, with loop
contributions included. There are several cases where higher-order
matrix elements are included at the Born level. Consider the case
of resonance production at a hadron collider, e.g.\ of a $\W$,
which is contained in the lowest-order process $\q \qbar' \to \W$.
In an inclusive description, additional jets recoiling against the
$\W$ may be generated by parton showers. {\Py} also contains
the two first-order processes $\q \g \to \W \q'$ and
$\q \qbar' \to \W \g$. The cross sections for these processes
are divergent when the $\pT \to 0$. In this region a correct
treatment would therefore have to take into account loop corrections,
which are not available in {\Py}.
Even without having these accessible, we know approximately what the
outcome should be. The virtual corrections have to cancel the
$\pT \to 0$ singularities of the real emission. The total cross
section of $\W$ production therefore receives finite
$\mathcal{O}(\alphas)$ corrections to the lowest-order answer. These
corrections can often be neglected to first approximation, except
when high precision is required. As for the shape of the $\W$ $\pT$
spectrum, the large cross section for low-$\pT$ emission has to be
interpreted as allowing more than one emission to take place. A
resummation procedure is therefore necessary to have matrix element
make sense at small $\pT$. The outcome is a cross section below the
na\"{\i}ve one, with a finite behaviour in the $\pT \to 0$ limit.
Depending on the physics application, one could then use {\Py} in one
of two ways. In an inclusive description, which is dominated by the
region of reasonably small $\pT$, the preferred option is
lowest-order matrix elements combined with parton showers, which
actually is one way of achieving the required resummation. For $\W$
production as background to some other process, say, only the
large-$\pT$ tail might be of interest. Then the shower approach may
be inefficient, since only few events will end up in the interesting
region, while the matrix-element alternative allows reasonable cuts to
be inserted from the beginning of the generation procedure. (One would
probably still want to add showers to describe additional softer
radiation, at the cost of some smearing of the original cuts.)
Furthermore, and not less importantly, the matrix elements should give
a more precise prediction of the high-$\pT$ event rate than the
approximate shower procedure.
In the particular case considered here, that of $\W$ production, and a
few similar processes, actually the shower has been improved by a matching
to first-order matrix elements, thus giving a decent description over the
whole $\pT$ range. This does not provide the first-order corrections to
the total $\W$ production rate, however, nor the possibility to select
only a high-$\pT$ tail of events.
\subsubsection{Parton showers}
The separation of radiation into initial- and final-state showers is
arbitrary, but very convenient. There are also situations where it
is appropriate: for instance, the process
$\ee \to \Z^0 \to \q \qbar$ only
contains final-state QCD radiation (QED radiation, however, is
possible both in the initial and final state), while
$\q \qbar \to \Z^0 \to \ee$ only contains initial-state QCD one.
Similarly, the distinction of emission as coming either from the
$\q$ or from the $\qbar$ is arbitrary. In general, the assignment
of radiation to a given mother parton is a good approximation
for an emission close to the direction of motion of that parton,
but not for the wide-angle emission in between two jets, where
interference terms are expected to be important.
In both initial- and final-state showers, the structure is given in
terms of branchings $a \to bc$, specifically $\e \to \e \gamma$,
$\q \to \q \g$, $\q \to \q \gamma$, $\g \to \g \g$, and
$\g \to \q \qbar$. (Further branchings, like $\gamma \to \e^+ \e^-$
and $\gamma \to \q \qbar$, could also have been added, but have not
yet been of interest.) Each of these processes is characterized by
a splitting kernel $P_{a \to bc}(z)$. The branching rate is
proportional to the integral $\int P_{a \to bc}(z) \, \d z$.
The $z$ value picked for a branching describes the energy sharing,
with daughter $b$ taking a fraction $z$ and daughter $c$ the
remaining $1-z$ of the mother energy. Once formed, the daughters
$b$ and $c$ may in turn branch, and so on.
Each parton is characterized by some virtuality scale $Q^2$,
which gives an approximate sense of time ordering to the cascade.
We stress here that somewhat different definition of $Q^2$ are
possible, and that {\Py} actually implements two distinct
alternatives, as you will see.
In the initial-state shower, $Q^2$ values are
gradually increasing as the hard scattering is approached, while
$Q^2$ is decreasing in the final-state showers.
Shower evolution is cut off at some lower
scale $Q_0$, typically around 1 GeV for QCD branchings. From above,
a maximum scale $Q_{\mmax}$ is introduced, where the showers are
matched to the hard interaction itself. The relation between
$Q_{\mmax}$ and the kinematics of the hard scattering
is uncertain, and the choice made can strongly affect the
amount of well-separated jets.
Despite a number of common traits, the initial- and final-state
radiation machineries are in fact quite different, and are described
separately below.
Final-state showers are time-like,
i.e.\ partons have $m^2 = E^2 - \mbf{p}^2 \geq 0$. The evolution
variable $Q^2$ of the cascade has therefore traditionally in {\Py}
been associated with the $m^2$ of the branching parton. As discussed
above, this choice is not unique, and in more recent versions of
{\Py}, a $\pT$-ordered shower algorithm, with $Q^2=\pTs=z(1-z)m^2$,
is available in
addition to the mass-ordered one. Regardless of the exact definition
of the ordering variable, the general strategy is the same:
starting from some maximum scale $Q^2_{\mmax}$, an original parton is evolved
downwards in $Q^2$ until a branching occurs. The selected $Q^2$ value defines
the mass of the branching parton, or the $\pT$ of the branching,
depending on whether the mass-ordering or the $\pT$-ordering is used.
In both cases, the $z$
value obtained from the splitting kernel represents the parton energy
division between the daughters. These daughters may now, in turn, evolve
downwards, in this case with maximum virtuality already defined by
the previous branching, and so on down to the $Q_0$ cut-off.
In QCD showers, corrections to the leading-log picture, so-called
coherence effects, lead to an ordering of subsequent emissions in terms
of decreasing angles. For the mass-ordering constraint, this does not
follow automatically, but is implemented as an additional
requirement on allowed emissions. The $\pT$-ordered shower leads to
the correct behaviour without such modifications \cite{Gus86}.
Photon emission is not affected by angular ordering. It is also
possible to obtain non-trivial correlations between azimuthal angles
in the various
branchings, some of which are implemented as
options. Finally, the theoretical analysis strongly
suggests the scale choice $\alphas = \alphas(\pT^2) =
\alphas(z(1-z)m^2)$, and this is the default in the program, for
\emph{both} shower algorithms.
The final-state radiation machinery is normally applied in the c.m.\
frame of the hard scattering or a decaying resonance. The total energy
and momentum of that subsystem is preserved, as is the direction of the
outgoing partons (in their common rest frame), where applicable.
In contrast to final-state showers, initial-state ones are space-like.
This means that, in the sequence of branchings $a \to bc$ that lead
up from the shower initiator to the hard interaction,
particles $a$ and $b$ have $m^2 = E^2 - \mbf{p}^2 <0$.
The `side branch' particle $c$, which does not participate
in the hard scattering, may be on the mass shell, or have a time-like
virtuality. In the latter case a time-like shower will evolve off
it, rather like the final-state radiation described above. To first
approximation, the evolution of the space-like main branch
is characterized by the
evolution variable $Q^2 = -m^2$, which is required to be strictly
increasing along the shower, i.e.\ $Q_b^2 > Q_a^2$. Corrections
to this picture have been calculated,
but are basically absent in {\Py}. Again, in more recent versions of
{\Py}, a $\pT$-ordered ISR algorithm is also available, with
$Q^2=\pTs=-(1-z)m^2$.
Initial-state radiation is handled within the backwards evolution
scheme. In this approach, the
choice of the hard scattering is based on the use of evolved
parton distributions, which means that
the inclusive effects of initial-state radiation are already
included. What remains is therefore to construct the exclusive
showers. This is done starting from the two incoming partons
at the hard interaction, tracing the showers `backwards in time',
back to the two shower initiators. In other words,
given a parton $b$, one tries to find the parton $a$ that branched
into $b$. The evolution in the Monte Carlo is therefore in
terms of a sequence of decreasing $Q^2$ (space-like virtuality or
transverse momentum, as applicable)
and increasing momentum fractions $x$. Branchings on
the two sides are interleaved in a common sequence of
decreasing $Q^2$ values.
In the above formalism, there is no real distinction between
gluon and photon emission. Some of the details actually do differ,
as will be explained in the full description.
The initial- and final-state radiation shifts around the kinematics of
the original hard interaction. In Deeply Inelastic Scattering, this
means that the $x$ and $Q^2$ values that can be derived from the
momentum of the scattered lepton do not automatically agree with the
values originally picked. In high-$\pT$ processes, it means that one no
longer has two jets with opposite and compensating $\pT$, but more
complicated topologies. Effects of any original kinematics selection
cuts are therefore smeared out, an unfortunate side-effect of the
parton-shower approach.
\subsection{Beam Remnants and Multiple Interactions}
\label{ss:overview:brmi}
To begin with, consider a hadron--hadron collision where only a single
parton--parton interaction occurs, i.e.\ we ignore the possibility of
multiple interactions for the moment. In that case,
the initial-state radiation algorithm
reconstructs one shower initiator in each beam. This initiator only
takes some fraction of the total beam energy, leaving behind a beam
remnant which takes the rest. For a proton beam, a $\u$ quark
initiator would leave behind a $\u \d$ diquark beam remnant, with an
antitriplet colour charge. The remnant is therefore colour-connected
to the hard interaction, and forms part of the same fragmenting
system. It is further customary to assign a primordial transverse
momentum to the shower initiator, to take into account the motion
of quarks inside the original hadron, at least as required by the
uncertainty principle by the proton size, probably augmented by
unresolved (i.e.\ not simulated) soft shower activity. This primordial
$k_{\perp}$ is selected according to some suitable distribution, and
the recoil is assumed to be taken up by the beam remnant.
Often the remnant is more complicated, e.g.\ a gluon initiator
would leave behind a $\u \u \d$ proton remnant system in a colour octet
state, which can conveniently be subdivided into a colour triplet
quark and a colour antitriplet diquark, each of which are
colour-connected to the hard interaction. The energy sharing between
these two remnant objects, and their relative transverse momentum,
introduces additional degrees of freedom, which are not understood
from first principles.
Na\"{\i}vely, one would expect an $\ep$ event to have only one beam
remnant, and an $\ee$ event none. This is not always correct, e.g.\
a $\gamma \gamma \to \q \qbar$ interaction in an $\ee$ event would
leave behind the $\e^+$ and $\e^-$ as beam remnants, and a
$\q \qbar \to \g \g$ interaction in resolved photoproduction in an
$\ee$ event would leave behind one $\e^{\pm}$ and one $\q$ or $\qbar$
in each remnant. Corresponding complications occur for
photoproduction in $\ep$ events.
There is another source of beam remnants. If parton distributions are
used to resolve an electron inside an electron, some of the original
energy is not used in the hard interaction, but is rather associated
with initial-state photon radiation. The initial-state shower is
in principle intended to trace this evolution and reconstruct the
original electron before any radiation at all took place. However,
because of cut-off procedures, some small amount may be left
unaccounted for.
Alternatively, you may have chosen to switch off initial-state
radiation altogether, but still preserved the resolved electron
parton distributions. In either case the remaining energy is given to
a single photon of vanishing transverse momentum, which is then
considered in the same spirit as `true' beam remnants.
So far we have assumed that
each event only contains one hard interaction, i.e.\ that each
incoming particle has only one parton which takes part in hard
processes, and that all other constituents sail through unaffected.
This is appropriate in $\ee$ or $\ep$ events, but not necessarily so in
hadron--hadron collisions (where a resolved photon counts as a hadron).
Here each of the beam particles contains a
multitude of partons, and so the probability for several interactions
in one and the same event need not be negligible. In principle these
additional interactions could arise because one single parton from
one beam scatters against several different partons from the other
%beam, or because several partons from each beam take place in
%referee: misprint
beam, or because several partons from each beam take part in
separate $2 \to 2$ scatterings. Both are expected, but combinatorics
should favour the latter, which is the mechanism considered in
{\Py}.
The dominant $2 \to 2$ QCD cross sections are
divergent for $\pT \to 0$, and drop rapidly for larger
$\pT$. Probably the lowest-order perturbative cross sections
will be regularized at small $\pT$ by colour coherence effects:
an exchanged gluon of small $\pT$ has a large transverse
wave function and can therefore not resolve the individual colour
charges of the two incoming hadrons; it will only couple to an
average colour charge that vanishes in the limit $\pT \to 0$.
In the program, some effective $\pTmin$ scale is therefore
introduced, below which the perturbative cross section is either
assumed completely vanishing or at least strongly damped.
Phenomenologically, $\pTmin$ comes out to be a number of
the order of 1.5--2.5~GeV, with some energy dependence.
In a typical `minimum-bias' event one therefore expects to find one
or a few scatterings at scales around or a bit above
$\pTmin$,
while a high-$\pT$ event also may have additional scatterings
at the $\pTmin$ scale. The probability to have several
high-$\pT$ scatterings in the same event is small, since the
cross section drops so rapidly with $\pT$.
The understanding of multiple interaction is still very primitive.
{\Py} therefore contains several different options.
These differ e.g.\
on the issue of the `pedestal' effect: is there an increased
probability or not for additional interactions in an event which
is known to contain a hard scattering, compared with one that
contains no hard interactions? Other differences concern the level
of detail in the generation of scatterings after the first one, and
the model that describes how the scatterings are intercorrelated in
flavour, colour, and momentum space.
The default underlying-event scenario obtained in a call to
\ttt{PYEVNT} corresponds to the so-called 'Tune A' \cite{Fie02}
(although with a slightly different energy dependence), which
reproduces many aspects of Tevatron data correctly. Starting from
{\Py} version 6.3, a more advanced model for the underlying event
is also available. This model is obtained by calling \ttt{PYEVNW}
instead of \ttt{PYEVNT}, but one should then not forget to also change
the relevant parameter settings to an appropriate tune of the new model.
\subsection{Hadronization}
QCD perturbation theory, formulated in terms of quarks and
gluons, is valid at short distances. At long distances, QCD
becomes strongly interacting and perturbation theory breaks
down. In this confinement regime, the coloured partons are
transformed into colourless hadrons, a process called either
hadronization or fragmentation. In this paper we reserve the
former term for the combination of fragmentation and the subsequent
decay of unstable particles.
The fragmentation process has yet to be understood from first
principles, starting from the QCD Lagrangian. This has left the
way clear for the development of a number of different
phenomenological models. Three main schools are usually
distinguished, string fragmentation (SF), independent
fragmentation (IF) and cluster fragmentation (CF),
but many variants and hybrids exist.
Being models, none of them can lay claims
to being `correct', although some may be better founded than
others. The best that can be aimed for is internal consistency,
a good representation of existing data, and a predictive power for
properties not yet studied or results at higher energies.
\subsubsection{String Fragmentation}
The original {\Je} program is intimately connected with string
fragmentation, in the form of the time-honoured `Lund model'. This
is the default for all {\Py} applications, but independent
fragmentation options also exist (although not actively maintained),
for applications where one wishes to study the
importance of string effects.
All current models are of a probabilistic and iterative nature.
This means that the fragmentation process as a whole is described in
terms of one or a few simple underlying branchings, of the type
jet $\to$ hadron + remainder-jet, string $\to$
hadron + remainder-string, and so on. At each branching,
probabilistic rules are given for the production of new flavours,
and for the sharing of energy and momentum between the products.
To understand fragmentation models, it is useful to start with
the simplest possible system, a colour-singlet $\q \qbar$ 2-jet
event, as produced in $\ee$ annihilation. Here lattice QCD studies
lend support to a linear confinement picture (in the absence
of dynamical quarks), i.e.\ the energy stored in the colour
dipole field between a charge and an anticharge increases linearly
with the separation between the charges, if the short-distance
Coulomb term is neglected. This is quite different
from the behaviour in QED, and is related to the presence of a
triple-gluon vertex in QCD. The details are not yet well
understood, however.
The assumption of linear confinement provides the starting point for
the string model. As the $\q$ and $\qbar$ partons move apart from
their common production vertex, the physical picture is that of a
colour flux tube (or maybe a colour vortex line) being stretched
between the $\q$ and the $\qbar$. The transverse dimensions
of the tube are of typical hadronic sizes, roughly 1 fm. If
the tube is assumed to be uniform along its length, this
automatically leads to a confinement picture with a linearly
rising potential. In order to obtain a Lorentz covariant and causal
description of the energy flow due to this linear confinement,
the most straightforward way is to use the dynamics of the massless
relativistic string with no transverse degrees of freedom.
The mathematical, one-dimensional string can
be thought of as parameterizing the position of the axis of a
cylindrically symmetric flux tube. From
hadron spectroscopy, the string constant, i.e.\ the amount of
energy per unit length, is deduced to be
$ \kappa \approx 1$ GeV/fm. The
expression `massless' relativistic string is somewhat of a
misnomer: $\kappa$ effectively corresponds to a `mass density' along
the string.
Let us now turn to the fragmentation process. As the $\q$ and
$\qbar$ move apart, the potential energy stored in the string
increases, and the string may break by the production of a new
$\q' \qbar'$ pair, so that the system splits into two
colour-singlet systems $\q \qbar'$ and $\q' \qbar$. If the invariant
mass of either of these string pieces is large enough, further
breaks may occur. In the Lund string model, the string break-up
process is assumed to proceed until only on-mass-shell hadrons
remain, each hadron corresponding to a small piece of string
with a quark in one end and an antiquark in the other.
In order to generate the quark--antiquark pairs $\q' \qbar'$ which
lead to string break-ups, the Lund model invokes the idea of
quantum mechanical tunnelling. This leads to a flavour-independent
Gaussian spectrum for the $\pT$ of $\q' \qbar'$ pairs.
Since the string is assumed to have no transverse excitations,
this $\pT$ is locally compensated between the quark and the
antiquark of the pair. The total $\pT$ of a hadron is made
up out of the $\pT$ contributions from the quark and
antiquark that together
form the hadron. Some contribution of very soft perturbative gluon
emission may also effectively be included in this description.
The tunnelling picture also implies a suppression of heavy-quark
production, $\u : \d : \s : \c \approx 1 : 1 : 0.3 : 10^{-11}$.
Charm and heavier quarks hence are not expected to be produced in
the soft fragmentation, but only in perturbative parton-shower
branchings $\g \to \q \qbar$.
When the quark and antiquark from two adjacent string breaks are
combined to form a meson, it is necessary to invoke an algorithm to
choose between the different allowed possibilities, notably
between pseudoscalar and vector mesons.
Here the string model is not particularly predictive. Qualitatively one
expects a $1 : 3$ ratio, from counting the number of spin states,
multiplied by some wave-function normalization factor, which should
disfavour heavier states.
A tunnelling mechanism can also be used to explain the production of
baryons. This is still a poorly understood area. In the simplest
possible approach, a diquark in a colour antitriplet state is just
treated like an ordinary antiquark, such that a string can break
either by quark--antiquark or antidiquark--diquark pair production.
A more complex scenario is the `popcorn' one, where
diquarks as such do not exist, but rather quark--antiquark pairs
are produced one after the other. This latter picture gives a less
strong correlation in flavour and momentum space between the
baryon and the antibaryon of a pair.
In general, the different string breaks are causally disconnected.
This means that it is possible to describe the breaks in any convenient
order, e.g.\ from the quark end inwards. One therefore is led to write
down an iterative scheme for the fragmentation, as follows.
Assume an initial quark $\q$ moving out along the $+z$ axis, with the
antiquark going out in the opposite direction.
By the production of a $\q_1 \qbar_1$ pair, a meson with flavour content
$\q \qbar_1$ is produced, leaving behind an unpaired quark $\q_1$.
A second pair $\q_2 \qbar_2$ may now be produced, to give a new meson
with flavours $\q_1 \qbar_2$, etc. At each step the produced
hadron takes some fraction of the available energy and momentum.
This process may be iterated until all energy is used up, with some
modifications close to the $\qbar$ end of the string in order to
make total energy and momentum come out right.
The choice of starting the fragmentation from the quark end is
arbitrary, however. A fragmentation process described in terms of
starting at the $\qbar$ end of the system and fragmenting towards
the $\q$ end should be equivalent.
This `left--right' symmetry constrains the allowed shape of the
fragmentation function $f(z)$, where $z$ is the fraction
of the remaining light-cone momentum $E \pm p_z$ (+ for the $\q$ jet,
$-$ for the $\qbar$ one) taken by each new particle.
The resulting `Lund symmetric fragmentation function' has two free
parameters, which are determined from data.
If several partons are moving apart from a common origin, the details
of the string drawing become more complicated. For a $\q \qbar \g$
event, a string is stretched from the
$\q$ end via the $\g$ to the $\qbar$ end, i.e.\
the gluon is a kink on the string, carrying energy and momentum.
As a consequence, the gluon has two string pieces attached, and
the ratio of gluon to quark string force is 2, a number which
can be compared with the ratio of colour charge Casimir operators,
$N_C/C_F = 2/(1-1/N_C^2) = 9/4$. In this, as in other
respects, the string model can be viewed as a variant of QCD
where the number of colours $N_C$ is not 3 but infinite.
Note that the factor 2 above does not depend on
the kinematical configuration: a smaller opening angle between
two partons corresponds to a smaller
string length drawn out per unit time, but also to an increased
transverse velocity of the string piece, which gives an exactly
compensating boost factor in the energy density per unit string
length.
The $\q \qbar \g$ string will fragment along its length. To first
approximation this means that there is
one fragmenting string piece between
$\q$ and $\g$ and a second one between $\g$ and $\qbar$. One hadron
is straddling both string pieces, i.e.\ sitting around the gluon
corner. The rest of the particles are produced as in two simple
$\q \qbar$ strings, but strings boosted with respect to the overall
c.m.\ frame. When considered in detail, the string motion and
fragmentation is more complicated, with the appearance of
additional string regions during the time evolution of the system.
These corrections are especially important for soft and
collinear gluons, since they provide a smooth transition between
events where such radiation took place and events where it did not.
Therefore the string fragmentation scheme is `infrared safe' with
respect to soft or collinear gluon emission.
Another possible colour topology arises when considering
baryon-number-violating processes, or events where more than one
valence quark has been knocked out of a beam baryon
(as can happen when multiple parton--parton interactions occur). In
this case, there will be three (anti-)colour carriers connected
antisymmetrically in colour, and of which no two may naturally be
considered to form a diquark system. The string topology will thus not
be of the simple $\q \qbar$ type, but rather a `Y' shaped topology is
spanned between the endpoints. The vertex of the `Y' topology comes to
be of special interest in the fragmentation, and will be referred to
as a `string junction'. Each of the three string pieces undergo a
fragmentation process subject to exactly the same principles as outlined
above, only a baryon containing the junction will eventually be formed.
The picture is essentially that of three jets going out, with the
junction baryon formed `in the middle', hence the junction baryon
will tend to have a soft spectrum when the jets are widely separated.
Note that, in the limit that two of the endpoints of the 'Y' come
close together, the diquark picture for beam remnants mentioned above
is effectively recovered, with only minor differences remaining.
For events that involve many partons, there may be several possible
topologies for their ordering along the string.
An example would be a $\q \qbar \g_1 \g_2$ (the gluon indices are here
used to label two different gluon-momentum vectors), where the
string can connect the partons in either of the sequences
$\q - \g_1 - \g_2 - \qbar$ and $\q - \g_2 - \g_1 - \qbar$.
The matrix elements that are calculable in perturbation theory
contain interference terms between these two possibilities, which
means that the colour flow is not always well-defined. Fortunately,
the interference terms are down in magnitude by a factor
$1/N_C^2$, where $N_C = 3$ is the number of colours, so
approximate recipes can be found. In the leading log shower
description, on the other hand, the rules for the colour flow are
well-defined.
A final comment: in the argumentation for the importance of colour flows
there is a tacit assumption that soft-gluon exchanges between partons
will not normally mess up the original colour assignment. Colour
rearrangement models provide toy scenarios wherein deviations from this
rule could be studied. Of particular interest has been the process
$\e^+ \e^- \to \W^+ \W^- \to \q_1 \qbar_2 \q_3 \qbar_4$, where the
original singlets $\q_1 \qbar_2$ and $\q_3 \qbar_4$ could be rearranged
to $\q_1 \qbar_4$ and $\q_3 \qbar_2$. So far, there are no experimental
evidence for dramatic effects of this kind, but the more realistic models
predict effects sufficiently small that these have not been ruled out.
Another example of nontrivial effects is that of Bose--Einstein correlations
between identical final-state particles, which reflect the true quantum
nature of the hadronization process.
\subsubsection{Decays}
A large fraction of the particles produced by fragmentation are
unstable and subsequently decay into the observable stable (or
almost stable) ones. It is therefore important to include all
particles with their proper mass distributions and decay properties.
Although involving little deep physics, this is less trivial than
it may sound: while a lot of experimental information is available,
there is also very much that is missing. For charm mesons,
it is necessary to put together measured exclusive branching ratios
with some inclusive multiplicity distributions to obtain a consistent
and reasonably complete set of decay channels, a rather delicate
task. For bottom even less is known, and for some $\B$ baryons only
a rather simple phase-space type of generator has been used for hadronic
decays.
Normally it is assumed that decay products are distributed according
to phase space, i.e.\ that there is no dynamics involved in their
relative distribution. However, in many cases additional assumptions
are necessary, e.g.\ for semileptonic decays of charm and bottom
hadrons one needs to include the proper weak matrix elements.
Particles may also be produced polarized and impart a non-isotropic
distribution to their decay products. Many of these effects are
not at all treated in the program. In fact, spin information is not
at all carried along, but has to be reconstructed explicitly when
needed.
This normal decay treatment makes use of a set of tables where
branching ratios and decay modes are stored. It encompasses all
hadrons made out of $\d$, $\u$, $\s$, $\c$ and $\b$ quarks, and also
the leptons. The decay products are hadrons, leptons and photons.
Some $\b\bbar$ states are sufficiently heavy that they are allowed to
decay to partonic states, like $\Upsilon \to \g\g\g$, which
subsequently fragment, but these are exceptions.
You may at will change the particle properties, decay channels or
%branching ratios of the above particles. There is no censorship what is
%referee: misprint
branching ratios of the above particles. There is no censorship of what is
allowed or not allowed, beyond energy--momentum, spin, and (electrical and
colour) charge conservation. There is also no impact e.g.\ on the cross
section of processes, since there is no way of knowing e.g.\ if the
restriction to one specific decay of a particle is because that decay
is of particular interest to us, or because recent measurement have
shown that this indeed is the only channel. Furthermore, the number of
particles produced of each species in the hadronization process is not
known beforehand, and so cannot be used to correctly bias the preceding
steps of the generation chain. All of this contrasts with the class of
`resonances' described above, in section \ref{sss:resdecintro}.
\clearpage
\section{Program Overview}
This section contains a diverse collection of information. The first
part is an overview of previous {\Je} and {\Py} versions.
The second gives instructions for installation of the program and
describes its philosophy: how it is constructed and how it is supposed
to be used. It also contains some information on how to read this manual.
The third and final part contains several examples of pieces
of code or short programs, to illustrate the general style of
program usage. This last part is mainly intended as an introduction for
completely new users, and can be skipped by more experienced ones.
The combined {\Py} package is completely self-contained.
Interfaces to externally defined subprocesses, parton-distribution
function libraries, SUSY parameter calculators, $\tau$ decay libraries,
and a time routine are provided, however, plus a few other optional
interfaces.
Many programs written by other persons make use of {\Py}, especially
the string fragmentation machinery. It is not the intention to give a
complete list here.
A majority of these programs are specific to given collaborations,
and therefore not publicly distributed. Below we give a list of a
few public programs from the `Lund group', which may have a
somewhat wider application. None of them are supported by the
{\Py} author team, so any requests should be directed to the persons
mentioned.
\begin{Itemize}
\item \tsc{Ariadne} is a generator for dipole emission, written
mainly by L. L\"onnblad \cite{Pet88}.
\item \tsc{LDCMC} is a related program for initial-state radiation
according to the Linked Dipole Chain model, also written mainly by
L.~L\"onnblad \cite{Kha99}.
\item \tsc{Aroma} is a generator for heavy-flavour processes in
leptoproduction, written by G.~Ingelman, J.~Rathsman and G. Schuler
\cite{Ing88}.
\item \tsc{Fritiof} is a generator for hadron--hadron, hadron--nucleus
and nucleus--nucleus collisions \cite{Nil87}.
\item \tsc{Lepto} is a leptoproduction event generator, written mainly
by G. Ingelman \cite{Ing80}. It can generate parton configurations
in Deeply Inelastic Scattering according to a number of possibilities.
\item \tsc{PomPyt} is a generator for pomeron interactions written by
G. Ingelman and collaborators \cite{Bru96}.
\end{Itemize}
One should also note that a version of {\Py} has been modified to
include the effects of longitudinally polarized incoming protons.
This is the work of St.~G\"ullenstern et al.\ \cite{Gul93}.
\subsection{Update History}
For the record, in Tables \ref{Jetsetver} and \ref{Pythiaver} we
list the official main versions of {\Je} and {\Py}, respectively,
with some brief comments.
\begin{table}[tp]
\caption{The main versions of {\Je}, with their date of
appearance, published manuals, and main changes from previous
versions. \protect\label{Jetsetver}}
\begin{center}
\begin{tabular}{|c|c|c|l|}
\hline
No. & Date & Publ. & Main new or improved
features \\[1mm]
\hline
1 & Nov 78 & \cite{Sjo78} & single-quark jets \\
2 & May 79 & \cite{Sjo79} & heavy-flavour jets \\
3.1 & Aug 79 & --- & 2-jets in $\ee$, preliminary 3-jets \\
3.2 & Apr 80 & \cite{Sjo80} & 3-jets in $\ee$ with full matrix
elements, \\
& & & toponium $\to \g \g \g$ decays \\
3.3 & Aug 80 & --- & softer fragmentation spectrum \\
4.1 & Apr 81 & --- & baryon production and diquark
fragmentation, \\
& & & fourth-generation quarks, larger
jet systems \\
4.2 & Nov 81 & --- & low-$\pT$ physics \\
4.3 & Mar 82 & \cite{Sjo82} & 4-jets and QFD structure in $\ee$, \\
& Jul 82 & \cite{Sjo83} & event-analysis routines \\
5.1 & Apr 83 & --- & improved string fragmentation scheme,
symmetric \\
& & & fragmentation, full 2$^{\mrm{nd}}$ order QCD
for $\ee$ \\
5.2 & Nov 83 & --- & momentum-conservation schemes for IF, \\
& & & initial-state photon radiation in $\ee$ \\
5.3 & May 84 & --- & `popcorn' model for baryon production \\
6.1 & Jan 85 & --- & common blocks restructured, parton showers \\
6.2 & Oct 85 & \cite{Sjo86} & error detection \\
6.3 & Oct 86 & \cite{Sjo87} & new parton-shower scheme \\
7.1 & Feb 89 & --- & new particle codes and common-block
structure, \\
& & & more mesons, improved decays, vertex
information, \\
& & & Abelian gluon model, Bose--Einstein
effects \\
7.2 & Nov 89 & --- & interface to new standard common block, \\
& & & photon emission in showers \\
7.3 & May 90 & \cite{Sjo92d} & expanded support for non-standard
particles \\
7.4 & Dec 93 & \cite{Sjo94} & updated particle data and defaults
\\[1mm]
\hline
\end{tabular}
\end{center}
\end{table}
\begin{table}[tp]
\caption{The main versions of {\Py}, with their date of
appearance, published manuals, and main changes from previous
versions. \protect\label{Pythiaver}}
\begin{center}
\begin{tabular}{|c|c|c|l|}
\hline
No. & Date & Publ. & Main new or improved features \\[1mm]
\hline
1 & Dec 82 & \cite{Ben84} & synthesis of
predecessors \tsc{Compton}, \tsc{Highpt} and \\
& & & \tsc{Kassandra} \\
2 & --- & & \\
3.1 & --- & & \\
3.2 & --- & & \\
3.3 & Feb 84 & \cite{Ben84a} & scale-breaking parton distributions \\
3.4 & Sep 84 & \cite{Ben85} & more efficient kinematics selection \\
4.1 & Dec 84 & & initial- and final-state parton showers,
$\W$ and $\Z$ \\
4.2 & Jun 85 & & multiple interactions \\
4.3 & Aug 85 & & $\W \W$, $\W \Z$, $\Z \Z$ and $\R$
processes \\
4.4 & Nov 85 & & $\gamma \W$, $\gamma \Z$, $\gamma \gamma$
processes \\
4.5 & Jan 86 & & $\H^0$ production, diffractive and
elastic events \\
4.6 & May 86 & & angular correlation in resonance
pair decays \\
4.7 & May 86 & & $\Z'^0$ and $\H^+$ processes \\
4.8 & Jan 87 & \cite{Ben87} & variable impact parameter
in multiple interactions \\
4.9 & May 87 & & $\g \H^+$ process \\
5.1 & May 87 & & massive matrix elements for heavy quarks \\
5.2 & Jun 87 & & intermediate boson scattering \\
5.3 & Oct 89 & & new particle and subprocess codes,
new common-block \\
& & & structure, new kinematics selection,
some \\
& & & lepton--hadron interactions, new subprocesses\\
5.4 & Jun 90 & & $s$-dependent widths, resonances not on the
mass shell, \\
& & & new processes, new parton distributions \\
5.5 & Jan 91 & & improved $\ee$ and $\ep$, several new
processes \\
5.6 & Sep 91 & \cite{Sjo92d} & reorganized parton distributions,
new processes, \\
& & & user-defined external processes \\
5.7 & Dec 93 & \cite{Sjo94} & new total cross sections,
photoproduction, top decay \\
6.1 & Mar 97 & \cite{Sjo01} &
merger with {\Je}, double precision, supersymmetry, \\
& & & technicolor, extra dimensions, etc. new processes, \\
& & & improved showers, virtual-photon processes \\
6.2 & Aug 01 & \cite{Sjo01a} & Les Houches Accord user processes,
$R$-parity violation \\
6.3 & Aug 03 & \cite{Sjo03a} & improved multiple interactions,
$\pT$--ordered showers,\\
& & & SUSY Les Houches Accord interface \\
6.4 & Mar 06 & this & none so far \\[1mm]
\hline
\end{tabular}
\end{center}
\end{table}
\paragraph{Versions before 6.1:}
All versions preceding {\Py}~6.1 should now be considered obsolete,
and are no longer maintained. For stable applications, the earlier
combination {\Je}~7.4 and {\Py}~5.7 could still be used, however.
\paragraph{Changes in version 6.1:}
The move from {\Je}~7.4 and {\Py}~5.7 to {\Py}~6.1 was a major one.
For reasons of space, individual points are therefore not listed
separately below, but only the main ones. The {\Py} web page contains
complete update notes, where all changes are documented by topic and
subversion.
The main new features of {\Py}~6.1, either present from the beginning
or added later on, include:
\begin{Itemize}
\item {\Py} and {\Je} have been merged.
\item All real variables are declared in double precision.
\item The internal mapping of particle codes has changed.
\item The supersymmetric process machinery of \tsc{SPythia} has been
included and further improved, with several new processes.
\item Many new processes of beyond-the-Standard-Model physics, in
areas such as technicolor and doubly-charged Higgs bosons.
\item An expanded description of QCD processes in virtual-photon
interactions, combined with a new machinery for the flux of virtual
photons from leptons.
\item Initial-state parton showers are matched to the next-to-leading
order matrix elements for gauge boson production.
\item Final-state parton showers are matched to a number of different
first-order matrix elements for gluon emission, including full
mass dependence.
\item The hadronization description of low-mass strings has been
improved, with consequences especially for heavy-flavour production.
\item An alternative baryon production model has been introduced.
\item Colour rearrangement is included as a new option, and several
alternative Bose-Einstein descriptions are added.
\end{Itemize}
\paragraph{Changes in version 6.2:}
By comparison, the move from {\Py}~6.1 to {\Py}~6.2 was rather less
dramatic. Again update notes tell the full story. Some of the main
new features, present from the beginning or added later on, which
may affect backwards compatibility, are:
\begin{Itemize}
\item A new machinery to handle user-defined external processes,
according to the Les Houches Accord standard in \cite{Boo01}. The old
machinery is no longer available. Some of the alternatives for the
\ttt{FRAME} argument in the \ttt{PYINIT} call have also been renamed
to make way for a new \ttt{'USER'} option.
\item The maximum size of the decay channel table has been increased
from 4000 to 8000, affecting the \ttt{MDME}, \ttt{BRAT} and \ttt{KFDP}
arrays in the \ttt{PYDAT3} common block.
\item A number of internally used and passed arrays, such as
\ttt{WDTP}, \ttt{WDTE}, \ttt{WDTPP}, \ttt{WDTEP}, \ttt{WDTPM},
\ttt{WDTEM}, \ttt{XLAM} and \ttt{IDLAM}, have been expanded
from dimension 300 to 400.
\item Lepton- and baryon-number-violating decay channels have been
included for supersymmetric particles \cite{Ska01, Sjo03}. Thus the
decay tables have grown considerably longer.
\item The string hadronization scheme has been improved and expanded
better to handle junction topologies, where three strings meet. This
is relevant for baryon-number-violating processes, and also for the
handling of baryon beam remnants. Thus new routines have been
introduced, and also e.g.\ new \ttt{K(I,1)} status codes.
\item A runtime interface to \tsc{Isasusy} has been added, for
determining the SUSY mass spectrum and mixing parameters more
accurately than with the internal {\Py} routines.
\item The Technicolor scenario is updated and extended. A new common
block, \ttt{PYTCSM}, is introduced for the parameters and switches
in Technicolor and related scenarios, and variables are moved to it
from a few other common blocks. New processes 381--388 are introduced
for standard QCD $2 \to 2$ interactions with Technicolor (or other
compositeness) extensions, while the processes 11, 12, 13, 28, 53,
68, 81 and 82 now revert back to being pure QCD.
\item The \ttt{PYSHOW} time-like showering routine has been expanded
to allow showering inside systems consisting of up to 80 particles,
which can be made use of in some resonance decays and in user-defined
processes.
\item The \ttt{PYSSPA} space-like showering routine has been expanded
with a $\q \to \q \gamma$ branching.
\item The \ttt{PYSIGH} routine has been split into several, in order to
make it more manageable. (It had increased to a size of over 7000
lines, which gave some compilers problems.) All the phase-space and
parton-density generic weights remain, whereas the process-specific
matrix elements have been grouped into new routines \ttt{PYSGQC},
\ttt{PYSGHF}, \ttt{PYSGWZ}, \ttt{PYSGHG}, \ttt{PYSGSU}, \ttt{PYSGTC}
and \ttt{PYSGEX}.
\item Some exotic particles and QCD effective states have been moved
from temporary flavour codes to a PDG-consistent naming, and a few new
codes have been introduced.
\item The maximum number of documentation lines in the beginning of the
event record has been expanded from 50 to 100.
\item The default parton distribution set for the proton is now
CTEQ 5L.
\item The default Standard Model Higgs mass has been changed to 115 GeV.
\end{Itemize}
\paragraph{Changes in version 6.3:}
The major changes in 6.3 was the introduction of
a new underlying--event framework for hadron collisions, together with
new transverse-momen\-tum-ordered initial- and final-state parton
showers. Other changes include
an interface to SUSY spectrum and decay calculators conformant to
the SUSY Les Houches Accord. Early releases of \Py~6.3, up to 6.312,
were still experimental, and did not include the new parton showers.
Details on these versions can be found in the update notes available
on the web.
We here concentrate on the structure of the program after (and including)
version 6.312.
Firstly, there are no incompatibilities
between \Py~6.2 and \Py~6.3 at the level of commonblock sizes or subroutine
arguments (except for the \ttt{PYSUGI} interface to \tsc{Isajet}, see below).
That is, any program that ran with {\Py}~6.2
also ought to run with {\Py}~6.3, without any change required.
Note, however, the following news:
\begin{Itemize}
\item The old (6.2) master event-generation routine \ttt{PYEVNT} is still
available; it uses the old underlying--event machinery (\ttt{PYMULT})
and parton showers (\ttt{PYSSPA} and \ttt{PYSHOW}). A `Tune A'--like set
of parameters \cite{Fie02} has been adopted as default. It reproduces
Tune A at the Tevatron, but implies a slower energy rescaling,
to give a more conservative estimate of the underlying activity at the LHC.
\item A new master event-generation routine \ttt{PYEVNW} has been introduced,
in parallel with \ttt{PYEVNT}. By calling \ttt{PYEVNW} instead of
\ttt{PYEVNT} a completely new scenario for multiple interactions \cite{Sjo04}
and parton showers \cite{Sjo04a} is obtained.
This framework has still only
been fully developed for hadron-hadron collisions (where a resolved
photon counts as a hadron). \emph{Note that many of
the default parameters now in \Py\ are directly
related to the \ttt{PYEVNT} model. These defaults are not appropriate to
use with
the new framework, and so all parameters should be set explicitly
whenever \ttt{PYEVNW} is used. See e.g.\ examples available on the web.}
\item The current \ttt{PYSUGI} run-time
interface to the \tsc{Isasusy} evolution package is
based on \tsc{Isajet} version 7.71 (from \Py\ 6.319 onwards).
The dimensions of the \ttt{SSPAR}, \ttt{SUGMG}, \ttt{SUGPAS} and
\ttt{SUGXIN} common blocks found in the \ttt{PYSUGI} routine could have
to be modified if another \tsc{Isajet} version is linked.
\item The new routine \ttt{PYSLHA} introduces a generic interface to
SUSY spectrum and decay calculators in agreement with the
SUSY Les Houches Accord \cite{Ska03} (SLHA), and thus offers a long-term
solution to the incompatibility problems noted above. Without SUSY switched
on, \ttt{PYSLHA} can still be used stand-alone to read in SLHA decay tables
for any particle, see section \ref{ss:parapartdat}.
\item From version 6.321 a run-time interface to FeynHiggs \cite{Hei99}
is available, to correct the Higgs sector of SUSY spectra obtained with
either of the \ttt{PYSUGI} and \ttt{PYSLHA} interfaces.
\item From version 6.321 a new set of toy models for investigating final
state colour reconnection effects are available, see \cite{San05}.
\item From version 6.322 an option exists to extend the particle
content recognized by \Py\ to that of the NMSSM, for use in the context
of interfaces to other programs, see \cite{Puk05}.
\item Some new processes added, notably $\Jpsi$ and $\Upsilon$ production
in an NRQCD colour-octet-model framework.
\end{Itemize}
The update notes tell exactly with what version a new feature or a
bug fix is introduced.
\paragraph{Changes in version 6.4:}
No new features are introduced in \Py\ 6.400 relative to the preceding
\Py\ 6.327. The main reason for the new version number is to
synchronize with the documentation, i.e.\ with this updated manual.
As changes are made to this baseline version, they will be documented
in the update notes.
\subsection{Program Installation}
\label{ss:install}
The {\Py} `master copy' is the one found on the web page\\[2mm]
\drawbox{\ttt{http://www.thep.lu.se/}$\sim$\ttt{torbjorn/Pythia.html}}\\
There you have, for several subversions \ttt{xxx}:
\begin{tabbing}
~~~ \= \ttt{pythia6xxx.f}~~~~~~~~~ \= the {\Py}~6.xxx code, \\
\> \ttt{pythia6xxx.tex} \> editions of this {\Py} manual, and \\
\> \ttt{pythia6xxx.update} \> plain text update notes to the manual.
\end{tabbing}
In addition to these, one may also find sample main programs and
other pieces of related software, and some physics papers.
The program is written essentially entirely in standard Fortran 77,
and should run on any platform with such a compiler. To a first
approximation, program compilation should therefore be straightforward.
Unfortunately, experience with many different compilers has been
uniform: the options available for obtaining optimized code may
actually produce erroneous code (e.g.\ operations inside \ttt{DO}
loops are moved out before them, where some of the variables have
not yet been properly set). Therefore the general advice is to
use a low optimization level. Note that this is often not the
default setting.
\ttt{SAVE} statements have been included in accordance with the
Fortran standard.
All default settings and particle and process data are stored in
\ttt{BLOCK DATA PYDATA}. This subprogram must be linked for a proper
functioning of the other routines. On some platforms this is not done
automatically but must be forced by you, e.g.\ by having a line
\begin{verbatim}
EXTERNAL PYDATA
\end{verbatim}
at the beginning of your main program. This applies in particular if
{\Py} is maintained as a library from which routines are to be loaded
only when they are needed. In this connection we note that the library
approach does not give any significant space advantages over a loading
of the packages as a whole, since a normal run will call on most of the
routines anyway, directly or indirectly.
With the move towards higher energies, e.g.\ for LHC applications,
single-precision (32 bit) real arithmetic has become inappropriate.
Therefore a declaration \ttt{IMPLICIT DOUBLE PRECISION(A-H,O-Z)} at
the beginning of each subprogram is inserted to ensure double-precision
(64 bit) real arithmetic. Remember that this also means that all calls
to {\Py} routines have to be done with real variables declared
correspondingly in the user-written calling program. An
\ttt{IMPLICIT INTEGER(I-N)} is also included to avoid
problems on some compilers. Integer functions beginning with \ttt{PY}
have to be declared explicitly. In total, therefore all routines
begin with
\begin{verbatim}
C...Double precision and integer declarations.
IMPLICIT DOUBLE PRECISION(A-H, O-Z)
IMPLICIT INTEGER(I-N)
INTEGER PYK,PYCHGE,PYCOMP
\end{verbatim}
and you are recommended to do the same in your main programs.
Note that, in running text and in description of common-block default
values, the more cumbersome double-precision notation is not always
made explicit, but code examples should be correct.
On a machine where \texttt{DOUBLE PRECISION} would give 128 bits,
it may make sense to use compiler options to revert to 64 bits,
since the program is anyway not constructed to make use of 128
bit precision.
Fortran~77 makes no provision for double-precision complex numbers.
Therefore complex numbers have been used only sparingly. However,
some matrix element expressions, mainly for supersymmetric and
technicolor processes, simplify considerably when written in terms
of complex variables. In order to achieve a uniform precision,
such variables have been declared \texttt{COMPLEX*16}, and are
manipulated with functions such as \texttt{DCMPLX} and
\texttt{DCONJG}. Affected are \texttt{PYSIGH}, \texttt{PYWIDT} and
several of the supersymmetry routines. Should the compiler not
accept this deviation from the standard, or some simple equivalent
thereof (like \texttt{DOUBLE COMPLEX} instead of \texttt{COMPLEX*16})
these code pieces could be rewritten to ordinary \texttt{COMPLEX},
also converting the real numbers involved to and from single precision,
with some drop in accuracy for the affected processes. \texttt{PYRESD}
already contains some ordinary \texttt{COMPLEX} variables, and should
not cause any problems.
Several compilers report problems when an odd number of integers
precede a double-precision variable in a common block. Therefore
an extra integer has been introduced as padding in a few instances,
e.g.\ \texttt{NPAD}, \texttt{MSELPD} and \texttt{NGENPD}.
Since Fortran~77 provides no date-and-time routine,
\texttt{PYTIME} allows a system-specific routine to be interfaced,
with some commented-out examples given in the code.
This routine is only used for cosmetic improvements of the output,
however, so can be left at the default with time 0 given.
A test program, \ttt{PYTEST}\label{p:PYTEST}, is included in the
{\Py} package. It is disguised as a subroutine, so you have to run a
main program
\begin{verbatim}
CALL PYTEST(1)
END
\end{verbatim}
This program will generate over a thousand events of different types,
under a variety of conditions. If {\Py} has not been properly
installed, this program is likely to crash, or at least generate a
number of erroneous events. This will then clearly be marked in the
output, which otherwise will just contain a few sample event listings
and a table of the number of different particles produced. To switch
off the output of normal events and final table, use \ttt{PYTEST(0)}
instead of \ttt{PYTEST(1)}. The final tally of errors detected should
read 0.
For a program written to run {\Py}~5 and {\Je}~7, most of the conversion
required for {\Py}~6 is fairly straightforward, and can be automatized.
Both a simple Fortran routine and a more sophisticated Perl \cite{Gar98}
script exist to this end, see the {\Py} web page. Some manual
checks and interventions may still be required.
\subsection{Program Philosophy}
The Monte Carlo program is built as a slave system, i.e.\ you,
the user, have to supply the main program. From this the various
subroutines are called on to execute specific tasks, after which
control is returned to the main program. Some of these tasks may
be very trivial, whereas the `high-level' routines by themselves
may make a large number of subroutine calls. Many routines are
not intended to be called directly by you, but only from
higher-level routines such as \ttt{PYEXEC}, \ttt{PYEEVT},
\ttt{PYINIT}, \ttt{PYEVNT}, or \ttt{PYEVNW}.
Basically, this means that there are three ways by which you
communicate with the
programs. First, by setting common-block variables, you specify
the details of how the programs should perform specific tasks,
e.g.\ which subprocesses should be generated, which
particle masses should be assumed, which coupling constants used,
which fragmentation scenarios, and so on with hundreds of options
and parameters. Second, by calling subroutines you tell the programs
to generate events according to the rules established above.
Normally there are few subroutine arguments, and those are usually
related to details of the physical situation, such as what
c.m.\ energy to assume for events. Third, you can either look at the
common block \ttt{PYJETS} to extract information on the generated
event, or you can call on various functions and subroutines
to analyse the event further for you.
It should be noted that, while the physics content is obviously at
the centre of attention, the {\Py} package
also contains a very extensive setup of auxiliary service routines.
The hope is that this will provide a comfortable
working environment, where not only events are generated, but where
you also linger on to perform a lot of the subsequent studies.
Of course, for detailed studies, it may be necessary to interface
the output directly to a detector simulation program.
The general rule is that all routines have names that are six
characters long, beginning with \ttt{PY}. Apart from dummy copies of
routines from other libraries, there are three exceptions
the length rules: \ttt{PYK}, \ttt{PYP} and \ttt{PYR}. The
former two functions are strongly coupled to the \ttt{K} and \ttt{P}
matrices in the \ttt{PYJETS} common block, while the latter is very
frequently used. Also common-block names are six characters long and
start with \ttt{PY}. There are three integer functions,
\ttt{PYK}, \ttt{PYCHGE} and \ttt{PYCOMP}. In all routines where
they are to be used, they have to be declared \ttt{INTEGER}.
%referee: mention appendices in the text.
An index to (almost) all subprograms and common-block variables
is found in Appendix B.
On the issue of initialization, the routines of different origin and
functionality behave quite differently. Routines that are intended to
be called from many different places, such as showers, fragmentation
and decays, require no specific initialization (except for the one
implied by the presence of \ttt{BLOCK DATA PYDATA}, see above), i.e.\
each event and each task stands on its own. Current common-block values
are used to perform the tasks in specific ways, and those rules can be
changed from one event to the next (or even within the generation of
one and the same event) without any penalty. The random-number generator
is initialized at the first call, but usually this is transparent.
In the core process generation machinery (e.g.\ selection of the hard
process kinematics), on the other hand, a sizable
amount of initialization is performed in the \ttt{PYINIT} call, and
thereafter the events generated by \ttt{PYEVNT} all obey the rules
established at that point. This improves the efficiency of the
generation process, and also ties in with the Monte Carlo integration
of the process cross section over many events. Therefore common-block
variables that specify methods and constraints to be used have to be
set before the \ttt{PYINIT} call and then not be changed afterwards,
with few exceptions. Of course, it is possible to perform several
\ttt{PYINIT} calls in the same run, but there is a significant time
overhead involved, so this is not something one would do for each new
event. The two older separate process generation routines \ttt{PYEEVT}
(and some of the routines called by it) and \ttt{PYONIA} also contain
some elements of initialization, where there are a few advantages if
events are generated in a coherent fashion. The cross section is not
as complicated here, however, so the penalty for reinitialization is
small, and also does not require any special user calls.
Apart from writing a title page, giving a brief initialization
information, printing error messages if need be,
and responding to explicit requests for listings, all tasks of the
program are performed `silently'. All output is directed to unit
\ttt{MSTU(11)}, by default 6, and it is up to you to set
this unit open for write. The only exceptions are
\ttt{PYRGET}, \ttt{PYRSET} and \ttt{PYUPDA} where, for obvious reasons,
the input/output file number is specified at each call. Here you
again have to see to it that proper read/write access is set.
The programs are extremely versatile, but the price to be paid for this
is having a large number of adjustable parameters and switches for
alternative modes of operation. No single user is ever likely to
need more than a fraction of the available options.
Since all these parameters and switches are assigned sensible default
values, there is no reason to worry about them until the need arises.
Unless explicitly stated (or obvious from the context) all switches and
parameters can be changed independently of each other. One should note,
however, that if only a few switches/parameters are changed, this may
result in an artificially bad agreement with data. Many disagreements
can often be cured by a subsequent retuning of some other parameters of
the model, in particular those that were once determined by
a comparison with data in the context of the default scenario.
For example, for $\ee$ annihilation, such a retuning could involve one
QCD parameter ($\alphas$ or $\Lambda$), the longitudinal fragmentation
function, and the average transverse momentum in fragmentation.
The program contains a number of checks that requested processes have
been implemented, that flavours specified for jet systems make sense,
that the energy is sufficient to allow hadronization, that the memory
space in \ttt{PYJETS} is large enough, etc. If anything goes wrong
that the program can catch (obviously this may not always be possible),
an error message will be printed and the treatment of the corresponding
event will be cut short. In serious cases, the program will abort.
As long as no error messages appear on the
output, it may not be worthwhile to look into the rules for error
checking, but if but one message appears, it should be enough cause for
alarm to receive prompt attention. Also warnings are sometimes printed.
These are less serious, and the experienced user might deliberately
do operations which go against the rules, but still can be made to
make sense in their context. Only the first few warnings will be
printed, thereafter the program will be quiet. By default, the program
is set to stop execution after ten errors, after printing
the last erroneous event.
It must be emphasized that not all errors will be caught. In particular,
one tricky question is what happens if an integer-valued common-block
switch or subroutine/function argument is used with a value that is not
defined. In some subroutine calls, a prompt return will be expedited,
but in most instances the subsequent action is entirely unpredictable,
and often completely haywire. The same goes for real-valued variables
that are assigned values outside the physically sensible range. One
example will suffice here: if \ttt{PARJ(2)} is defined as the
$\s / \u$ suppression factor, a value $>1$ will not give more
profuse production of $\s$ than of $\u$, but actually a spillover
into $\c$ production. Users, beware!
\subsection{Manual Conventions}
In the manual parts of this report, some conventions are used.
All names of subprograms, common blocks and variables are given in
upper-case `typewriter' style, e.g.\ \ttt{MSTP(111) = 0}. Also
program examples are given in this style.
If a common-block variable must have a value set at the beginning of
execution, then a default value is stored in the block data
subprogram \ttt{PYDATA}. Such a default value is
usually indicated by a `(D = \ldots)' immediately after the variable
name, e.g.
\begin{entry}
\iteme{MSTJ(1) :} (D = 1) choice of fragmentation scheme.
\end{entry}
All variables in the fragmentation-related common blocks (with very few
exceptions, clearly marked) can be freely changed from one event to the
next, or even within the treatment of one single event; see discussion
on initialization in the previous section. In the process-generation
machinery common blocks the situation is more complicated. The values of
many switches and parameters are used already in the \ttt{PYINIT} call,
and cannot be changed after that. The problem is mentioned in the
preamble to the afflicted common blocks, which in particular means
\ttt{PYPARS} and \ttt{PYSUBS}. For the variables which may still
be changed from one event to the next, a `(C)' is added after
the `(D = \ldots)' statement.
Normally, variables internal to the program are kept in separate
common blocks and arrays, but in a few cases such internal variables
appear among arrays of switches and parameters, mainly for historical
reasons. These are denoted by `(R)' for variables you may want to
read, because they contain potentially interesting information, and
by `(I)' for purely internal variables. In neither case may the
variables be changed by you.
In the description of a switch, the alternatives that this
switch may take are often enumerated, e.g.
\begin{entry}
\iteme{MSTJ(1) :} (D = 1) choice of fragmentation scheme.
\begin{subentry}
\iteme{= 0 :} no jet fragmentation at all.
\iteme{= 1 :} string fragmentation according to the Lund model.
\iteme{= 2 :} independent fragmentation, according to specification
in \ttt{MSTJ(2)} and \ttt{MSTJ(3)}.
\end{subentry}
\end{entry}
If you then use any value other than 0, 1 or 2, results are
undefined. The action could even be different in different parts
of the program, depending on the order in which the alternatives are
identified.
It is also up to you to choose physically sensible values for
parameters: there is no check on the allowed ranges of variables.
We gave an example of this at the end of the preceding section.
Subroutines you are expected to use are enclosed in a box at the
point where they are defined:
\drawbox{CALL PYLIST(MLIST)}
\boxsep
This is followed by a description of input or output parameters.
The difference between input and output is not explicitly marked,
but should be obvious from the context. In fact, the event-analysis
routines of section \ref{ss:evanrout} return values,
while all the rest only have input variables.
Routines that are only used internally are not boxed in.
However, we use boxes for all common blocks, so as to
enhance the readability.
In running text, often specific switches and parameters will be
mentioned, without a reference to the place where they are described
further. The Index at the very end of the document allows you to
find this place. Often names for switches begin with \ttt{MST} and
parameters with \ttt{PAR}. No common-block variables begin with
\ttt{PY}. There is thus no possibility to confuse an array element
with a function or subroutine call.
An almost complete list of common blocks, with brief comments
on their main functions, is the following:
\begin{verbatim}
C...The event record.
COMMON/PYJETS/N,NPAD,K(4000,5),P(4000,5),V(4000,5)
C...Parameters.
COMMON/PYDAT1/MSTU(200),PARU(200),MSTJ(200),PARJ(200)
C...Particle properties + some flavour parameters.
COMMON/PYDAT2KCHG(500,4),PMAS(500,4),PARF(2000),VCKM(4,4)
C...Decay information.
COMMON/PYDAT3/MDCY(500,3),MDME(8000,2),BRAT(8000),KFDP(8000,5)
C...Particle names
COMMON/PYDAT4/CHAF(500,2)
CHARACTER CHAF*16
C...Random number generator information.
COMMON/PYDATR/MRPY(6),RRPY(100)
C...Selection of hard scattering subprocesses.
COMMON/PYSUBS/MSEL,MSELPD,MSUB(500),KFIN(2,-40:40),CKIN(200)
C...Parameters.
COMMON/PYPARS/MSTP(200),PARP(200),MSTI(200),PARI(200)
C...Internal variables.
COMMON/PYINT1/MINT(400),VINT(400)
C...Process information.
COMMON/PYINT2/ISET(500),KFPR(500,2),COEF(500,20),ICOL(40,4,2)
C...Parton distributions and cross sections.
COMMON/PYINT3/XSFX(2,-40:40),ISIG(1000,3),SIGH(1000)
C...Resonance width and secondary decay treatment.
COMMON/PYINT4/MWID(500),WIDS(500,5)
C...Generation and cross section statistics.
COMMON/PYINT5/NGENPD,NGEN(0:500,3),XSEC(0:500,3)
C...Process names.
COMMON/PYINT6/PROC(0:500)
CHARACTER PROC*28
C...Total cross sections.
COMMON/PYINT7/SIGT(0:6,0:6,0:5)
C...Photon parton distributions: total and valence only.
COMMON/PYINT8/XPVMD(-6:6),XPANL(-6:6),XPANH(-6:6),XPBEH(-6:6),
&XPDIR(-6:6)
COMMON/PYINT9/VXPVMD(-6:6),VXPANL(-6:6),VXPANH(-6:6),VXPDGM(-6:6)
C...Colour tag information in the Les Houches Accord format.
COMMON/PYCTAG/NCT, MCT(4000,2)
C...Partons and their scales for the new pT-ordered final-state showers.
PARAMETER (MAXNUP=500)
COMMON/PYPART/NPART,NPARTD,IPART(MAXNUP),PTPART(MAXNUP)
C...Multiple interactions in the new model.
COMMON/PYINTM/KFIVAL(2,3),NMI(2),IMI(2,800,2),NVC(2,-6:6),
&XASSOC(2,-6:6,240),XPSVC(-6:6,-1:240),PVCTOT(2,-1:1),
&XMI(2,240),Q2MI(240),IMISEP(0:240)
C...Hardest initial-state radiation in the new model.
COMMON/PYISMX/MIMX,JSMX,KFLAMX,KFLCMX,KFBEAM(2),NISGEN(2,240),
&PT2MX,PT2AMX,ZMX,RM2CMX,Q2BMX,PHIMX
C...Possible joined interactions in backwards evolution.
COMMON/PYISJN/MJN1MX,MJN2MX,MJOIND(2,240)
C...Supersymmetry parameters.
COMMON/PYMSSM/IMSS(0:99),RMSS(0:99)
C...Supersymmetry mixing matrices.
COMMON/PYSSMT/ZMIX(4,4),UMIX(2,2),VMIX(2,2),SMZ(4),SMW(2),
&SFMIX(16,4),ZMIXI(4,4),UMIXI(2,2),VMIXI(2,2)
C...R-parity-violating couplings in supersymmetry.
COMMON/PYMSRV/RVLAM(3,3,3), RVLAMP(3,3,3), RVLAMB(3,3,3)
C...Internal parameters for R-parity-violating processes.
COMMON/PYRVNV/AB(2,16,2),RMS(0:3),RES(6,5),IDR,IDR2,DCMASS,KFR(3)
COMMON/PYRVPM/RM(0:3),A(2),B(2),RESM(2),RESW(2),MFLAG
LOGICAL MFLAG
C...Parameters for Gauss integration of supersymmetric widths.
COMMON/PYINTS/XXM(20)
COMMON/PYG2DX/X1
C...Parameters of TechniColor Strawman Model and other compositeness.
COMMON/PYTCSM/ITCM(0:99),RTCM(0:99)
C...Histogram information.
COMMON/PYBINS/IHIST(4),INDX(1000),BIN(20000)
C...HEPEVT common block.
PARAMETER (NMXHEP=4000)
COMMON/HEPEVT/NEVHEP,NHEP,ISTHEP(NMXHEP),IDHEP(NMXHEP),
&JMOHEP(2,NMXHEP),JDAHEP(2,NMXHEP),PHEP(5,NMXHEP),VHEP(4,NMXHEP)
DOUBLE PRECISION PHEP,VHEP
C...User process initialization common block.
INTEGER MAXPUP
PARAMETER (MAXPUP=100)
INTEGER IDBMUP,PDFGUP,PDFSUP,IDWTUP,NPRUP,LPRUP
DOUBLE PRECISION EBMUP,XSECUP,XERRUP,XMAXUP
COMMON/HEPRUP/IDBMUP(2),EBMUP(2),PDFGUP(2),PDFSUP(2),
&IDWTUP,NPRUP,XSECUP(MAXPUP),XERRUP(MAXPUP),XMAXUP(MAXPUP),
&LPRUP(MAXPUP)
C...User process event common block.
INTEGER MAXNUP
PARAMETER (MAXNUP=500)
INTEGER NUP,IDPRUP,IDUP,ISTUP,MOTHUP,ICOLUP
DOUBLE PRECISION XWGTUP,SCALUP,AQEDUP,AQCDUP,PUP,VTIMUP,SPINUP
COMMON/HEPEUP/NUP,IDPRUP,XWGTUP,SCALUP,AQEDUP,AQCDUP,IDUP(MAXNUP),
&ISTUP(MAXNUP),MOTHUP(2,MAXNUP),ICOLUP(2,MAXNUP),PUP(5,MAXNUP),
&VTIMUP(MAXNUP),SPINUP(MAXNUP)
\end{verbatim}
\subsection{Getting Started with the Simple Routines}
\label{ss:JETstarted}
Normally {\Py} is expected to take care of the full event generation
process. At times, however, one may want to access the more simple
underlying routines, which allow a large flexibility to `do it
yourself'. We therefore start with a few cases of this kind,
at the same time introducing some of the more frequently used
utility routines.
As a first example, assume that you want to study the production of
$\u \ubar$ 2-jet systems at 20 GeV energy. To do this, write a main
program
\begin{verbatim}
IMPLICIT DOUBLE PRECISION(A-H, O-Z)
CALL PY2ENT(0,2,-2,20D0)
CALL PYLIST(1)
END
\end{verbatim}
and run this program, linked together with {\Py}. The routine
\ttt{PY2ENT}
is specifically intended for storing two entries (partons or particles).
The first argument (0) is a command to perform fragmentation and decay
directly after the entries have been stored, the second and third that
the two entries are $\u$ (2) and $\ubar$ ($-2$), and the last that the
c.m.\ energy of the pair is 20 GeV, in double precision. When this is run,
the resulting event is stored in the \ttt{PYJETS} common block. This
information can then be read out by you. No output is produced by
\ttt{PY2ENT} itself, except for a title page which appears once for
every {\Py} run.
Instead the second command, to \ttt{PYLIST}, provides a simple visible
summary of the information stored in \ttt{PYJETS}. The argument (1)
indicates that the short version should be used, which is suitable
for viewing the listing directly on an 80-column terminal screen.
It might look as shown here.
\begin{verbatim}
Event listing (summary)
I particle/jet KS KF orig p_x p_y p_z E m
1 (u) A 12 2 0 0.000 0.000 10.000 10.000 0.006
2 (ubar) V 11 -2 0 0.000 0.000 -10.000 10.000 0.006
3 (string) 11 92 1 0.000 0.000 0.000 20.000 20.000
4 (rho+) 11 213 3 0.098 -0.154 2.710 2.856 0.885
5 (rho-) 11 -213 3 -0.227 0.145 6.538 6.590 0.781
6 pi+ 1 211 3 0.125 -0.266 0.097 0.339 0.140
7 (Sigma0) 11 3212 3 -0.254 0.034 -1.397 1.855 1.193
8 (K*+) 11 323 3 -0.124 0.709 -2.753 2.968 0.846
9 p~- 1 -2212 3 0.395 -0.614 -3.806 3.988 0.938
10 pi- 1 -211 3 -0.013 0.146 -1.389 1.403 0.140
11 pi+ 1 211 4 0.109 -0.456 2.164 2.218 0.140
12 (pi0) 11 111 4 -0.011 0.301 0.546 0.638 0.135
13 pi- 1 -211 5 0.089 0.343 2.089 2.124 0.140
14 (pi0) 11 111 5 -0.316 -0.197 4.449 4.467 0.135
15 (Lambda0) 11 3122 7 -0.208 0.014 -1.403 1.804 1.116
16 gamma 1 22 7 -0.046 0.020 0.006 0.050 0.000
17 K+ 1 321 8 -0.084 0.299 -2.139 2.217 0.494
18 (pi0) 11 111 8 -0.040 0.410 -0.614 0.751 0.135
19 gamma 1 22 12 0.059 0.146 0.224 0.274 0.000
20 gamma 1 22 12 -0.070 0.155 0.322 0.364 0.000
21 gamma 1 22 14 -0.322 -0.162 4.027 4.043 0.000
22 gamma 1 22 14 0.006 -0.035 0.422 0.423 0.000
23 p+ 1 2212 15 -0.178 0.033 -1.343 1.649 0.938
24 pi- 1 -211 15 -0.030 -0.018 -0.059 0.156 0.140
25 gamma 1 22 18 -0.006 0.384 -0.585 0.699 0.000
26 gamma 1 22 18 -0.034 0.026 -0.029 0.052 0.000
sum: 0.00 0.000 0.000 0.000 20.000 20.000
\end{verbatim}
(A few blanks have been removed between the columns to make it fit into
the format of this text.) Look in the particle/jet column and note
that the first two lines are the original $\u$ and $\ubar$.
The parentheses enclosing the names, `\ttt{(u)}' and `\ttt{(ubar)}',
are there as a reminder that these partons actually have been allowed
to fragment. The partons are still retained so that event histories
can be studied. Also note that the \ttt{KF} (flavour code) column
contains 2 in the first line and $-2$ in the second. These are the
codes actually stored to denote the presence of a $\u$ and a $\ubar$,
cf.\ the \ttt{PY2ENT} call, while the names written are just
conveniences used when producing visible output. The \ttt{A} and
\ttt{V} near the end of the particle/jet column indicate the beginning
and end of a string (or cluster, or independent fragmentation) parton
system; any intermediate entries belonging to the same system would
have had an \ttt{I} in that column. (This gives a poor man's
representation of an up-down arrow, $\updownarrow$.)
In the \ttt{orig} (origin) column, the zeros indicate that $\u$ and
$\ubar$ are two initial entries. The subsequent line, number 3, denotes
the fragmenting $\u \ubar$ string system as a whole, and has origin 1,
since the first parton of this string system is entry number 1. The
particles in lines 4--10 have origin 3 to denote that they come
directly from the fragmentation of this string. In string
fragmentation it is not meaningful to say that a particle comes from
only the $\u$ quark or only the $\ubar$ one. It is the string system
as a whole that gives a $\rho^+$, a $\rho^-$, a $\pi^+$, a
$\Sigma^0$, a $\K^{*+}$, a $\pbar^-$, and a $\pi^-$. Note that some
of the particle names are again enclosed in parentheses, indicating
that these particles are not present in the final state either, but have
decayed further. Thus the $\pi^-$ in line 13 and the $\pi^0$ in line
14 have origin 5, as an indication that they come from the decay of the
$\rho^-$ in line 5. Only the names not enclosed in parentheses
remain at the end of the fragmentation/decay chain, and
are thus experimentally observable. The actual status code used to
distinguish between different classes of entries is given in the
\ttt{KS} column; codes in the range 1--10 correspond to remaining
entries, and those above 10 to those that have fragmented or decayed.
The columns with \ttt{p\_x}, \ttt{p\_y}, \ttt{p\_z}, \ttt{E} and
\ttt{m} are quite self-explanatory. All momenta, energies and
masses are given in units of GeV, since the speed of light
is taken to be $c = 1$.
Note that energy and momentum are conserved at each step of the
fragmentation/decay process (although there exist options where this is
not true). Also note that the $z$ axis plays the r\^ole of preferred
direction, along which the original partons are placed.
The final line is
intended as a quick check that nothing funny happened. It contains the
summed charge, summed momentum, summed energy and invariant mass of the
final entries at the end of the fragmentation/decay chain, and the
values should agree with the input implied by the \ttt{PY2ENT}
arguments. (In fact, warnings would normally appear on the output if
anything untoward happened, but that is another story.)
The above example has illustrated roughly what information is to be had
in the event record, but not so much about how it is stored. This is
better seen by using a 132-column format for listing events. Try e.g.
the following program
\begin{verbatim}
IMPLICIT DOUBLE PRECISION(A-H, O-Z)
CALL PY3ENT(0,1,21,-1,30D0,0.9D0,0.7D0)
CALL PYLIST(2)
CALL PYEDIT(3)
CALL PYLIST(2)
END
\end{verbatim}
where a 3-jet $\d \g \dbar$ event is generated in the first executable
line and listed in the second. This listing will contain the numbers as
directly stored in the common block \ttt{PYJETS}
\begin{verbatim}
COMMON/PYJETS/N,NPAD,K(4000,5),P(4000,5),V(4000,5)
\end{verbatim}
For particle \ttt{I}, \ttt{K(I,1)} thus gives information on whether
or not a parton or particle has fragmented or decayed, \ttt{K(I,2)}
gives the particle code, \ttt{K(I,3)} its origin, \ttt{K(I,4)} and
\ttt{K(I,5)} the position of fragmentation/decay products, and
\ttt{P(I,1})--\ttt{P(I,5)} momentum, energy and mass. The number
of lines in current use is given by \ttt{N}, i.e.\
1 $\leq$ \ttt{I} $\leq$ \ttt{N}. The \ttt{V} matrix contains decay
vertices; to view those \ttt{PYLIST(3)} has to be used. \ttt{NPAD}
is a dummy, needed to avoid some compiler troubles. It is important
to learn the rules for how information is stored in \ttt{PYJETS}.
The third executable line in the program illustrates another important
point about {\Py}: a number of routines are available for manipulating
and analyzing the event record after the event has been generated.
Thus \ttt{PYEDIT(3)} will remove everything except stable charged
particles, as shown by the result of the second \ttt{PYLIST} call.
More advanced possibilities include things like sphericity or
clustering routines. {\Py} also contains some simple routines for
histogramming, used to give self-contained examples of analysis
procedures.
Apart from the input arguments of subroutine calls, control on the
doings of {\Py} may be imposed via many common blocks. Here
sensible default values are always provided. A user might want to
switch off all particle decays by putting \ttt{MSTJ(21) = 0} or
increase the $\s / \u$ ratio in fragmentation by putting
\ttt{PARJ(2) = 0.40D0}, to give but two examples. It is by exploring
the possibilities offered here that {\Py} can be turned into an
extremely versatile tool, even if all the nice physics is already
present in the default values.
As a final, semi-realistic example, assume that the $\pT$
spectrum of $\pi^+$ particles is to be studied in 91.2 GeV
$\ee$ annihilation events, where $\pT$ is to be defined with
respect to the sphericity axis. Using the internal routines for
simple histogramming, a complete program might look like
\begin{verbatim}
C...Double precision and integer declarations.
IMPLICIT DOUBLE PRECISION(A-H, O-Z)
IMPLICIT INTEGER(I-N)
INTEGER PYK,PYCHGE,PYCOMP
C...Common blocks.
COMMON/PYJETS/N,NPAD,K(4000,5),P(4000,5),V(4000,5)
C...Book histograms.
CALL PYBOOK(1,'pT spectrum of pi+',100,0D0,5D0)
C...Number of events to generate. Loop over events.
NEVT=100
DO 110 IEVT=1,NEVT
C...Generate event. List first one.
CALL PYEEVT(0,91.2D0)
IF(IEVT.EQ.1) CALL PYLIST(1)
C...Find sphericity axis.
CALL PYSPHE(SPH,APL)
C...Rotate event so that sphericity axis is along z axis.
CALL PYEDIT(31)
C...Loop over all particles, but skip if not pi+.
DO 100 I=1,N
IF(K(I,2).NE.211) GOTO 100
C...Calculate pT and fill in histogram.
PT=SQRT(P(I,1)**2+P(I,2)**2)
CALL PYFILL(1,PT,1D0)
C...End of particle and event loops.
100 CONTINUE
110 CONTINUE
C...Normalize histogram properly and list it.
CALL PYFACT(1,20D0/NEVT)
CALL PYHIST
END
\end{verbatim}
Study this program, try to understand what happens at each step, and
run it to check that it works. You should then be ready to look
at the relevant sections of this report and start writing your own
programs.
\subsection{Getting Started with the Event Generation Machinery}
\label{ss:PYTstarted}
A run with the full {\Py} event generation machinery has to be more
strictly organized than the simple examples above, in that it is
necessary to initialize the
generation before events can be generated, and in that it is
not possible to change switches and parameters freely during
the course of the run. A fairly precise recipe for how a run
should be structured can therefore be given.
Thus, the nowadays normal usage of {\Py} can be subdivided into
three steps.
\begin{Enumerate}
\item The initialization step. It is here that all the basic
characteristics of the coming generation are specified.
The material in this section includes the following.
\begin{Itemize}
\item Declarations for double precision and integer variables and
integer functions:
\begin{verbatim}
IMPLICIT DOUBLE PRECISION(A-H, O-Z)
IMPLICIT INTEGER(I-N)
INTEGER PYK,PYCHGE,PYCOMP
\end{verbatim}
\vspace{-\baselineskip}
\item Common blocks, at least the following, and maybe some more:
\begin{verbatim}
COMMON/PYJETS/N,NPAD,K(4000,5),P(4000,5),V(4000,5)
COMMON/PYDAT1/MSTU(200),PARU(200),MSTJ(200),PARJ(200)
COMMON/PYSUBS/MSEL,MSELPD,MSUB(500),KFIN(2,-40:40),CKIN(200)
COMMON/PYPARS/MSTP(200),PARP(200),MSTI(200),PARI(200)
\end{verbatim}
\vspace{-\baselineskip}
\item Selection of required processes. Some fixed `menus'
of subprocesses can be selected with different \ttt{MSEL}
values, but with {\tt MSEL}=0 it is possible to compose
`\`a la carte', using the subprocess numbers.
To generate processes 14, 18 and 29, for instance, one needs
\begin{verbatim}
MSEL=0
MSUB(14)=1
MSUB(18)=1
MSUB(29)=1
\end{verbatim}
\vspace{-\baselineskip}
\item Selection of kinematics cuts in the \ttt{CKIN} array.
To generate hard scatterings with 50~GeV $\leq \pT \leq$
100~GeV, for instance, use
\begin{verbatim}
CKIN(3)=50D0
CKIN(4)=100D0
\end{verbatim}
\vspace{-\baselineskip}
Unfortunately, initial- and final-state radiation will shift
around the kinematics of the hard scattering, making the effects
of cuts less predictable. One therefore always has to be very
careful that no desired event configurations are cut out.
\item Definition of underlying physics scenario, e.g.\ Higgs mass.
\item Selection of parton-distribution sets, $Q^2$ definitions,
showering and multiple interactions parameters,
and all other details of the generation.
\item Switching off of generator parts not needed for toy
simulations, e.g.\ fragmentation for parton level studies.
\item Initialization of the event generation procedure. Here
kinematics is set up, maxima of differential cross sections
are found for future Monte Carlo generation, and a number of
other preparatory tasks carried out. Initialization is performed
by \ttt{PYINIT}, which should be called only after the switches
and parameters above have been set to their desired values. The
frame, the beam particles and the energy have to be specified,
e.g.
\begin{verbatim}
CALL PYINIT('CMS','p','pbar',1800D0)
\end{verbatim}
\vspace{-\baselineskip}
\item Any other initial material required by you, e.g.
histogram booking.
\end{Itemize}
\item The generation loop. It is here that events are generated
and studied. It includes the following tasks:
\begin{Itemize}
\item Generation of the next event, with
\begin{verbatim}
CALL PYEVNT
\end{verbatim}
\vspace{-\baselineskip}
or, for the new multiple interactions and showering model,
\begin{verbatim}
CALL PYEVNW
\end{verbatim}
\vspace{-\baselineskip}
\item Printing of a few events, to check that everything is
working as planned, with
\begin{verbatim}
CALL PYLIST(1)
\end{verbatim}
\vspace{-\baselineskip}
\item An analysis of the event for properties of interest,
either directly reading out information from the
\ttt{PYJETS} common block
or making use of the utility routines in {\Py}.
\item Saving of events on disk or tape, or interfacing to detector
simulation.
\end{Itemize}
\item The finishing step. Here the tasks are:
\begin{Itemize}
\item Printing a table of deduced cross sections, obtained as a
by-product of the Monte Carlo generation activity, with the
command
\begin{verbatim}
CALL PYSTAT(1)
\end{verbatim}
\vspace{-\baselineskip}
\item Printing histograms and other user output.
\end{Itemize}
\end{Enumerate}
To illustrate this structure, imagine a toy example, where one wants
to simulate the production of a 300 GeV Higgs particle. In
{\Py}, a program for this might look something like the
following.
\begin{verbatim}
C...Double precision and integer declarations.
IMPLICIT DOUBLE PRECISION(A-H, O-Z)
IMPLICIT INTEGER(I-N)
INTEGER PYK,PYCHGE,PYCOMP
C...Common blocks.
COMMON/PYJETS/N,NPAD,K(4000,5),P(4000,5),V(4000,5)
COMMON/PYDAT1/MSTU(200),PARU(200),MSTJ(200),PARJ(200)
COMMON/PYDAT2/KCHG(500,4),PMAS(500,4),PARF(2000),VCKM(4,4)
COMMON/PYDAT3/MDCY(500,3),MDME(8000,2),BRAT(8000),KFDP(8000,5)
COMMON/PYSUBS/MSEL,MSELPD,MSUB(500),KFIN(2,-40:40),CKIN(200)
COMMON/PYPARS/MSTP(200),PARP(200),MSTI(200),PARI(200)
C...Number of events to generate. Switch on proper processes.
NEV=1000
MSEL=0
MSUB(102)=1
MSUB(123)=1
MSUB(124)=1
C...Select Higgs mass and kinematics cuts in mass.
PMAS(25,1)=300D0
CKIN(1)=290D0
CKIN(2)=310D0
C...For simulation of hard process only: cut out unnecessary tasks.
MSTP(61)=0
MSTP(71)=0
MSTP(81)=0
MSTP(111)=0
C...Initialize and list partial widths.
CALL PYINIT('CMS','p','p',14000D0)
CALL PYSTAT(2)
C...Book histogram.
CALL PYBOOK(1,'Higgs mass',50,275D0,325D0)
C...Generate events. Look at first few.
DO 200 IEV=1,NEV
CALL PYEVNT
IF(IEV.LE.3) CALL PYLIST(1)
C...Loop over particles to find Higgs and histogram its mass.
DO 100 I=1,N
IF(K(I,1).LT.20.AND.K(I,2).EQ.25) HMASS=P(I,5)
100 CONTINUE
CALL PYFILL(1,HMASS,1D0)
200 CONTINUE
C...Print cross sections and histograms.
CALL PYSTAT(1)
CALL PYHIST
END
\end{verbatim}
Here 102, 123 and 124 are the three main Higgs production
graphs $\g \g \rightarrow \hrm$, $\Z \Z \rightarrow \hrm$, and
$\W \W \rightarrow \hrm$, and \ttt{MSUB(ISUB) = 1} is the command to
switch on process {\ISUB}. Full freedom to combine subprocesses
`\`a la carte' is ensured by \ttt{MSEL = 0}; ready-made `menus'
can be ordered with other \ttt{MSEL} numbers. The \ttt{PMAS}
command sets the mass of the Higgs, and the \ttt{CKIN}
variables the desired mass range of the Higgs --- a Higgs with a
300 GeV nominal mass actually has a fairly broad Breit--Wigner type
mass distribution. The \ttt{MSTP} switches that come next are there to
modify the generation procedure, in this case to switch off initial-
and final-state radiation, multiple interactions among beam jets,
and fragmentation, to give only the `parton skeleton' of the hard
process. The \ttt{PYINIT} call initializes {\Py}, by finding
maxima of cross sections, recalculating the Higgs decay properties
(which depend on the Higgs mass), etc. The decay properties can
be listed with \ttt{PYSTAT(2)}.
Inside the event loop, \ttt{PYEVNT}
is called to generate an event, and \ttt{PYLIST(1)} to list the event.
The information used by \ttt{PYLIST(1)} is the event record, stored
in the common block \ttt{PYJETS}. Here one finds all produced particles,
both final and intermediate ones, with information on particle
species and event history (\ttt{K} array), particle momenta
(\ttt{P} array) and production vertices (\ttt{V} array).
In the loop over all particles produced, \ttt{1} through \ttt{N},
the Higgs particle is found by its code, \ttt{K(I,2) = 25},
and its mass is stored in \ttt{P(I,5)}.
After all events have been generated, \ttt{PYSTAT(1)}
gives a summary of the number of events generated in the
various allowed channels, and the inferred cross sections.
In the run above, a typical event listing might look like the following.
\begin{verbatim}
Event listing (summary)
I particle/jet KF p_x p_y p_z E m
1 !p+! 2212 0.000 0.000 8000.000 8000.000 0.938
2 !p+! 2212 0.000 0.000-8000.000 8000.000 0.938
======================================================================
3 !g! 21 -0.505 -0.229 28.553 28.558 0.000
4 !g! 21 0.224 0.041 -788.073 788.073 0.000
5 !g! 21 -0.505 -0.229 28.553 28.558 0.000
6 !g! 21 0.224 0.041 -788.073 788.073 0.000
7 !H0! 25 -0.281 -0.188 -759.520 816.631 300.027
8 !W+! 24 120.648 35.239 -397.843 424.829 80.023
9 !W-! -24 -120.929 -35.426 -361.677 391.801 82.579
10 !e+! -11 12.922 -4.760 -160.940 161.528 0.001
11 !nu_e! 12 107.726 39.999 -236.903 263.302 0.000
12 !s! 3 -62.423 7.195 -256.713 264.292 0.199
13 !cbar! -4 -58.506 -42.621 -104.963 127.509 1.350
======================================================================
14 (H0) 25 -0.281 -0.188 -759.520 816.631 300.027
15 (W+) 24 120.648 35.239 -397.843 424.829 80.023
16 (W-) -24 -120.929 -35.426 -361.677 391.801 82.579
17 e+ -11 12.922 -4.760 -160.940 161.528 0.001
18 nu_e 12 107.726 39.999 -236.903 263.302 0.000
19 s A 3 -62.423 7.195 -256.713 264.292 0.199
20 cbar V -4 -58.506 -42.621 -104.963 127.509 1.350
21 ud_1 A 2103 -0.101 0.176 7971.328 7971.328 0.771
22 d V 1 -0.316 0.001 -87.390 87.390 0.010
23 u A 2 0.606 0.052 -0.751 0.967 0.006
24 uu_1 V 2203 0.092 -0.042-7123.668 7123.668 0.771
======================================================================
sum: 2.00 0.00 0.00 0.00 15999.98 15999.98
\end{verbatim}
The above event listing is abnormally short, in part because some
columns of information were removed to make it fit into this text,
in part because all initial- and final-state QCD radiation, all
non-trivial beam jet structure, and all fragmentation was inhibited
in the generation. Therefore only the skeleton of the process is
visible. In lines 1 and 2 one recognizes the two incoming protons.
In lines 3 and 4 are incoming partons before initial-state radiation
and in 5 and 6 after --- since there is no such radiation they coincide
here. Line 7 shows the Higgs produced by $\g \g$ fusion, 8 and 9 its
decay products and 10--13 the second-step decay products. Up to this
point lines give a summary of the event history, indicated by the
exclamation marks that surround particle names (and also reflected in
the \ttt{K(I,1)} code, not shown). From line 14 onwards come the
particles actually produced in the final states, first in lines 14--16
particles that subsequently decayed, which have their names surrounded
by brackets, and finally the particles and partons left in the end,
including beam remnants.
Here this also includes a number of unfragmented partons, since
fragmentation was inhibited. Ordinarily, the listing would have gone
on for a few hundred more lines, with the particles produced in the
fragmentation and their decay products. The final line gives total
charge and momentum, as a convenient check that nothing unexpected
happened. The first column of the listing
is just a counter, the second gives the particle name and information
on status and string drawing (the \ttt{A} and \ttt{V}), the third
the particle-flavour code (which is used to give the name),
and the subsequent columns give the momentum components.
One of the main problems is to select kinematics efficiently. Imagine
for instance that one is interested in the production of a single
$\Z$ with a
transverse momentum in excess of 50 GeV. If one tries to generate
the inclusive sample of $\Z$ events, by the basic production graphs
$\q \qbar \rightarrow \Z$, then most events will have low transverse
momenta and will have to be discarded. That any of the desired events
are produced at all is due to the initial-state generation machinery,
which can build up transverse momenta for the incoming
$\q$ and $\qbar$. However, the amount
of initial-state radiation cannot be constrained beforehand. To
increase the efficiency, one may therefore turn to the higher-order
processes $\q \g \rightarrow \Z \q$ and $\q \qbar \rightarrow \Z \g$,
where
already the hard subprocess gives a transverse momentum to the $\Z$.
This transverse momentum can be constrained as one wishes, but again
initial- and final-state radiation will smear the picture. If one were
to set a $\pT$ cut at 50 GeV for the hard-process generation,
those events where the $\Z$ was given only 40 GeV in the hard process
but got the rest from initial-state radiation would be missed.
Not only therefore would cross sections
come out wrong, but so might the typical event shapes. In the end,
it is therefore necessary to find some reasonable compromise, by
starting the generation at 30 GeV, say, if one knows that only rarely
do events below this value fluctuate up to 50 GeV. Of course, most
events will therefore not contain a $\Z$ above 50 GeV, and one will
have to live with some inefficiency. It is not uncommon that only one
event out of ten can be used, and occasionally it can be even worse.
If it is difficult to set kinematics, it is often easier to set the
flavour content of a process. In a Higgs study, one might wish, for
example, to consider the decay $\hrm^0 \rightarrow \Z^0 \Z^0$, with each
$\Z^0 \rightarrow \ee$ or $\mu^+ \mu^-$. It is therefore necessary to
inhibit all other $\hrm^0$ and $\Z^0$ decay channels, and also to adjust
cross sections to take into account this change, all of which is fairly
straightforward. The same cannot be said for decays of ordinary hadrons,
where the number produced in a process is not known beforehand, and
therefore inconsistencies easily can arise if one tries to force
specific decay channels.
In the examples given above, all run-specific parameters are set in
the code (in the main program; alternatively it could be in a
subroutine called by the main program). This approach is allowing
maximum flexibility to change parameters during the course of the run.
However, in many experimental collaborations one does not want to allow
this freedom, but only one set of parameters, to be read in from an
external file at the beginning of a run and thereafter never changed.
This in particular applies when {\Py} is to be linked with other
libraries, such as \tsc{Geant} \cite{Bru89} and detector-specific software.
While a linking of a normal-sized main program with {\Py} is
essentially instantaneous on current platforms (typically less than a
second), this may not hold for other libraries. For this purpose one
then needs a parser of {\Py} parameters, the core of which can be
provided by the \ttt{PYGIVE} routine.
As an example, consider a main program of the form
\begin{verbatim}
C...Double precision and integer declarations.
IMPLICIT DOUBLE PRECISION(A-H, O-Z)
IMPLICIT INTEGER(I-N)
INTEGER PYK,PYCHGE,PYCOMP
C...Input and output strings.
CHARACTER FRAME*12,BEAM*12,TARGET*12,PARAM*100
C...Read parameters for PYINIT call.
READ(*,*) FRAME,BEAM,TARGET,ENERGY
C...Read number of events to generate, and to print.
READ(*,*) NEV,NPRT
C...Loop over reading and setting parameters/switches.
100 READ(*,'(A)',END=200) PARAM
CALL PYGIVE(PARAM)
GOTO 100
C...Initialize PYTHIA.
200 CALL PYINIT(FRAME,BEAM,TARGET,ENERGY)
C...Event generation loop
DO 300 IEV=1,NEV
CALL PYEVNT
IF(IEV.LE.NPRT) CALL PYLIST(1)
300 CONTINUE
C...Print cross sections.
CALL PYSTAT(1)
END
\end{verbatim}
and a file \ttt{indata} with the contents
\begin{verbatim}
CMS,p,p,14000.
1000,3
! below follows commands sent to PYGIVE
MSEL=0 ! Mix processes freely
MSUB(102)=1 ! g + g -> h0
MSUB(123)=1 ! Z0 + Z0 -> h0
MSUB(124)=1 ! W+ + W- -> h0
PMAS(25,1)=300. ! Higgs mass
CKIN(1)=290. ! lower cutoff on mass
CKIN(2)=310. ! upper cutoff on mass
MSTP(61)=0 ! no initial-state showers
MSTP(71)=0 ! no final-state showers
MSTP(81)=0 ! no multiple interactions
MSTP(111)=0 ! no hadronization
\end{verbatim}
Here the text following the exclamation marks is interpreted
as a comment by \ttt{PYGIVE}, and thus purely intended to
allow better documentation of changes. The main program could
then be linked to {\Py}, to an executable \ttt{a.out},
and run e.g.\ with a Unix command line
\begin{verbatim}
a.out < indata > output
\end{verbatim}
to produce results on the file \ttt{output}. Here the \ttt{indata}
could be changed without requiring a recompilation. Of course, the
main program would have to be more realistic, e.g.\ with events
saved to disk or tape, but the principle should be clear.
Further examples of how to use \Py\ are available on the
{\Py} webpage.
\clearpage
\section{Monte Carlo Techniques}
Quantum mechanics introduces a concept of randomness in the behaviour
of physical processes. The virtue of event generators is that this
randomness can be simulated by the use of Monte Carlo techniques.
In the process, the program authors have to use some ingenuity to
find the most efficient way to simulate an assumed probability
distribution. A detailed description of possible techniques would
carry us too far, but
in this section some of the most frequently used approaches are
presented, since they will appear in discussions in
subsequent sections. Further examples may be found e.g.\ in
\cite{Jam80}.
First of all one assumes the existence of a random number generator.
This is a (Fortran) function which, each time it is called, returns
a number $R$ in the range between 0 and 1, such that the inclusive
distribution of numbers $R$ is flat in the range,
and such that different numbers $R$ are uncorrelated. The random
number generator that comes with {\Py} is described at the
end of this section, and we defer the discussion until then.
\subsection{Selection From a Distribution}
\label{ss:MCdistsel}
The situation that is probably most common is that we know a
function $f(x)$ which is non-negative in the allowed $x$ range
$x_{\mmin} \leq x \leq x_{\mmax}$. We want to select an $x$
`at random' so that the probability in a small interval $\d x$
around a given $x$ is proportional to $f(x) \, \d x$. Here $f(x)$
might be a fragmentation function, a differential cross section,
or any of a number of distributions.
One does not
have to assume that the integral of $f(x)$ is explicitly normalized
to unity: by the Monte Carlo procedure of picking exactly one
accepted $x$ value, normalization is implicit in the final result.
Sometimes the integral of $f(x)$ does carry a physics content of
its own, as part of an overall weight factor
we want to keep track of. Consider, for instance, the case when $x$
represents one or several phase-space variables and $f(x)$ a
differential cross section; here the integral has a meaning of total
cross section for the process studied. The task of a Monte Carlo is
then, on the one hand, to generate events one at a time, and, on the
other hand, to estimate the total cross section.
The discussion of this important example is
deferred to section \ref{ss:PYTcrosscalc}.
If it is possible to find a primitive function $F(x)$ which
has a known inverse $F^{-1}(x)$, an $x$ can be found as
follows (method 1):
\begin{eqnarray}
& \displaystyle{ \int_{x_{\mmin}}^{x} f(x) \, \d x =
R \int_{x_{\mmin}}^{x_{\mmax}} f(x) \, \d x } & \nonumber \\[1mm]
\Longrightarrow & x = F^{-1}(F(x_{\mmin}) +
R(F(x_{\mmax}) - F(x_{\mmin}))) ~. &
\end{eqnarray}
The statement of the first line is that a fraction $R$ of the
total area under $f(x)$ should be to the left of $x$.
However, seldom are functions of interest so nice that the
method above works. It is therefore necessary to use more complicated
schemes.
Special tricks can sometimes be found. Consider e.g.\ the generation
of a Gaussian $f(x) = \exp(-x^2)$. This function is not integrable,
but if we combine it with the same Gaussian distribution of a second
variable $y$, it is possible to transform to polar coordinates
\begin{equation}
f(x) \, \d x \, f(y) \, \d y = \exp(-x^2-y^2) \, \d x \, \d y =
r \exp(-r^2) \, \d r \, \d \varphi ~,
\end{equation}
and now the $r$ and $\varphi$ distributions may be easily generated
and recombined to yield $x$. At the same time we get a second number
$y$, which can also be used. For the generation of transverse momenta
in fragmentation, this is very convenient, since in fact we want to
assign two transverse degrees of freedom.
If the maximum of $f(x)$ is known, $f(x) \leq f_{\mmax}$ in the
$x$ range considered, a hit-or-miss method will always yield the
correct answer (method 2):
\begin{Enumerate}
\item select an $x$ with even probability in the allowed range, i.e.\
$x = x_{\mmin} + R(x_{\mmax} - x_{\mmin})$;
\item compare a (new) $R$ with the ratio $f(x)/f_{\mmax}$;
if $f(x)/f_{\mmax} \le R$, then reject the $x$ value
and return to point 1 for a new try;
\item otherwise the most recent $x$ value is retained as final answer.
\end{Enumerate}
The probability that $f(x)/f_{\mmax} > R$ is proportional to
$f(x)$; hence the correct distribution of retained $x$ values.
The efficiency of this method, i.e.\ the average probability that
an $x$ will be retained, is $(\int \, f(x) \, \d x)
/ (f_{\mmax}(x_{\mmax} - x_{\mmin}))$.
The method is acceptable if this number is not too low, i.e.\ if
$f(x)$ does not fluctuate too wildly.
Very often $f(x)$ does have narrow spikes, and it may not even be
possible to define an $f_{\mmax}$. An example of the former
phenomenon is a
function with a singularity just outside the allowed region,
an example of the latter an integrable singularity just at the
$x_{\mmin}$ and/or $x_{\mmax}$ borders.
Variable transformations may then be used to make a function
smoother. Thus a function $f(x)$ which blows up as $1/x$ for
$x \rightarrow 0$, with an $x_{\mmin}$ close to 0, would instead
be roughly constant if transformed to the variable $y = \ln x$.
The variable transformation strategy may be seen as a combination
of methods 1 and 2, as follows. Assume the existence of a function
$g(x)$, with $f(x) \leq g(x)$ over the $x$ range of interest.
Here $g(x)$ is picked to be a `simple' function, such that the
primitive function $G(x)$ and its inverse $G^{-1}(x)$ are known.
Then (method 3):
\begin{Enumerate}
\item select an $x$ according to the distribution $g(x)$, using
method 1;
\item compare a (new) $R$ with the ratio $f(x)/g(x)$;
if $f(x)/g(x) \le R$, then reject the $x$ value
and return to point 1 for a new try;
\item otherwise the most recent $x$ value is retained as final answer.
\end{Enumerate}
This works, since the first step will select $x$ with a probability
$g(x) \, \d x = \d G(x)$ and the second retain this choice with
probability $f(x)/g(x)$. The total probability to pick a value $x$
is then just the product of the two, i.e.\ $f(x) \, \d x$.
If $f(x)$ has several spikes, method 3 may work for each spike
separately, but it may not be possible to find a $g(x)$ that
covers all of them at the same time, and which still has an
invertible primitive function. However, assume that
we can find a function $g(x) = \sum_i g_i(x)$,
such that $f(x) \leq g(x)$ over
the $x$ range considered, and such that the functions $g_i(x)$
each are non-negative and simple, in the sense that we can find
primitive functions and their inverses. In that case (method 4):
\begin{Enumerate}
\item select an $i$ at random, with relative probability given
by the integrals
\begin{equation}
\int_{x_{\mmin}}^{x_{\mmax}} g_i(x) \, \d x =
G_i(x_{\mmax}) - G_i(x_{\mmin}) ~; \nonumber
\end{equation}
\item for the $i$ selected, use method 1 to find an $x$, i.e.\
\begin{equation}
x = G_i^{-1}(G_i(x_{\mmin}) +
R(G_i(x_{\mmax})-G_i(x_{\mmin}))) ~;
\nonumber
\end{equation}
\item compare a (new) $R$ with the ratio $f(x)/g(x)$;
if $f(x)/g(x) \le R$, then reject the $x$ value
and return to point 1 for a new try;
\item otherwise the most recent $x$ value is retained as final answer.
\end{Enumerate}
This is just a trivial extension of method 3, where steps 1 and 2
ensure that, on the average, each $x$ value picked there is
distributed according to $g(x)$: the first step picks $i$
with relative probability $\int g_i(x) \, \d x$, the second $x$ with
absolute probability $g_i(x) / \int g_i(x) \, \d x$ (this is one place
where one must remember to do normalization correctly); the
product of the two is therefore $g_i(x)$ and the sum over all $i$
gives back $g(x)$.
We have now arrived at an approach that is sufficiently powerful for
a large selection of problems. In general,
for a function $f(x)$ which is known to have sharp peaks in a few
different places, the generic behaviour at each peak separately
may be covered by one or a few simple functions $g_i(x)$, to which
one adds a few more $g_i(x)$ to cover the basic behaviour away
from the peaks. By a suitable selection
of the relative strengths of the different $g_i$'s, it is
possible to find a function $g(x)$ that matches well the general
behaviour of $f(x)$, and thus achieve a reasonable Monte Carlo
efficiency.
The major additional complication is when $x$ is a multidimensional
variable. Usually the problem is not so much $f(x)$ itself, but
rather that the phase-space boundaries may be very complicated.
If the boundaries factorize it is possible to pick phase-space
points restricted to the desired region. Otherwise the region may
have to be inscribed in a hyper-rectangle, with points picked within
the whole hyper-rectangle but only retained if they are inside the
allowed region. This may lead to a significant loss in efficiency.
Variable transformations may often make the allowed region easier to
handle.
There are two main methods to handle several dimensions, each with its
set of variations. The first method is based on a factorized ansatz,
i.e.\ one attempts to find a function $g(\mbf{x})$ which is
everywhere larger than $f(\mbf{x})$, and which can be factorized
into $g(\mbf{x}) = g^{(1)}(x_1) \, g^{(2)}(x_2) \cdots g^{(n)}(x_n)$,
where $\mbf{x} = (x_1,x_2,\ldots,x_n)$. Here each $g^{(j)}(x_j)$ may
in its turn be a sum of functions $g^{(j)}_i$, as in method 4 above.
First, each $x_j$ is selected independently, and afterwards the ratio
$f(\mbf{x})/g(\mbf{x})$ is used to determine whether to retain the
point.
The second method is useful if the boundaries of the allowed region
can be written in a form where the maximum range of $x_1$ is known,
the allowed range of $x_2$ only depends on $x_1$, that of $x_3$ only
on $x_1$ and $x_2$, and so on until $x_n$, whose range may depend on
all the
preceding variables. In that case it may be possible to find a function
$g(\mbf{x})$ that can be integrated over $x_2$ through $x_n$ to yield
a simple function of $x_1$, according to which $x_1$ is selected.
Having done that, $x_2$ is selected according to a distribution
which now depends on $x_1$, but with $x_3$ through $x_n$ integrated
over. In particular, the allowed range for $x_2$ is known.
The procedure is continued until $x_n$ is reached, where now the
function depends on all the preceding $x_j$ values. In the end, the
ratio $f(\mbf{x})/g(\mbf{x})$ is again used to determine whether to
retain the point.
\subsection{The Veto Algorithm}
\label{ss:vetoalg}
The `radioactive decay' type of problems is very common, in particular
in parton showers, but it is also used, e.g.\ in the multiple
interactions description in {\Py}. In this kind of problems
there is one variable $t$, which may be thought of as giving a kind
of time axis along which different events are ordered. The
probability that `something will happen' (a nucleus decay, a
parton branch) at time $t$ is described by a function $f(t)$, which
is non-negative in the range of $t$ values to be studied. However,
this na\"{\i}ve probability is modified by the additional requirement
that something can only happen at time $t$ if it did not happen
at earlier times $t' < t$. (The original nucleus cannot decay once
again if it already did decay; possibly the decay products may decay
in their turn, but that is another question.)
The probability that nothing has happened by time $t$ is expressed by
the function ${\cal N}(t)$ and the differential probability that
something happens at time $t$ by ${\cal P}(t)$. The basic equation
then is
\begin{equation}
{\cal P}(t) = - \frac{\d {\cal N}}{\d t} = f(t) \, {\cal N}(t) ~.
\end{equation}
For simplicity, we shall assume that the process starts at time
$t = 0$, with ${\cal N}(0) = 1$.
The above equation can be solved easily if one notes that
$\d {\cal N} / {\cal N} = \d \ln {\cal N}$:
\begin{equation}
{\cal N}(t) = {\cal N}(0) \exp \left\{ - \int_0^t f(t') \, \d t'
\right\}
= \exp \left\{ - \int_0^t f(t') \, \d t' \right\} ~,
\end{equation}
and thus
\begin{equation}
{\cal P}(t) = f(t) \exp \left\{ - \int_0^t f(t') \, \d t' \right\} ~.
\label{mc:Pveto}
\end{equation}
With $f(t) = c$ this is nothing but the textbook formulae for
radioactive decay. In particular, at small times the correct
decay probability, ${\cal P}(t)$, agrees well with the input
one, $f(t)$, since the exponential factor is close to unity there.
At larger $t$, the exponential gives a dampening which ensures that
the integral of ${\cal P}(t)$ never can exceed unity, even if the
integral of $f(t)$ does. The exponential can be seen as the
probability that nothing happens between the original time 0 and
the final time $t$. In the parton-shower language, this corresponds
to the so-called Sudakov form factor.
If $f(t)$ has a primitive function with a known inverse, it is easy
to select $t$ values correctly:
\begin{equation}
\int_0^t {\cal P}(t') \, \d t' = {\cal N}(0) - {\cal N}(t) =
1 - \exp \left\{ - \int_0^t f(t') \, \d t' \right\} = 1 - R ~,
\end{equation}
which has the solution
\begin{equation}
F(0) - F(t) = \ln R ~~~ \Longrightarrow ~~~
t = F^{-1}(F(0) - \ln R) ~.
\end{equation}
If $f(t)$ is not sufficiently nice, one may again try to find a
better function $g(t)$, with $f(t) \leq g(t)$ for all $t \geq 0$.
However to use method 3 with this $g(t)$ would not work, since the
method would not correctly take
into account the effects of the exponential term in ${\cal P}(t)$.
Instead one may use the so-called veto algorithm:
\begin{Enumerate}
\item start with $i = 0$ and $t_0 = 0$;
\item add 1 to $i$ and select $t_i = G^{-1}(G(t_{i-1}) - \ln R)$,
i.e.\ according to $g(t)$, but with the constraint that
$t_i > t_{i-1}$,
\item compare a (new) $R$ with the ratio $f(t_i)/g(t_i)$;
if $f(t_i)/g(t_i) \le R$, then return to point 2 for a new try;
\item otherwise $t_{i}$ is retained as final answer.
\end{Enumerate}
It may not be apparent why this works. Consider, however, the various
ways in which one can select a specific time $t$. The probability that
the first try works, $t = t_1$, i.e.\ that no intermediate $t$ values
need be rejected, is given by
\begin{equation}
{\cal P}_0 (t) = \exp \left\{ - \int_0^t g(t') \, \d t' \right\}
\, g(t) \, \frac{f(t)}{g(t)}
= f(t) \exp \left\{ - \int_0^t g(t') \, \d t' \right\} ~,
\end{equation}
where the exponential times $g(t)$ comes from eq.~(\ref{mc:Pveto})
applied to $g$, and the ratio $f(t)/g(t)$ is the probability that
$t$ is
accepted. Now consider the case where one intermediate time $t_1$ is
rejected and $t = t_2$ is only accepted in the second step.
This gives
\begin{equation}
{\cal P}_1 (t) = \int_0^t \d t_1
\exp \left\{ - \int_0^{t_1} g(t') \, \d t' \right\} g(t_1)
\left[ 1 - \frac{f(t_1)}{g(t_1)} \right]
\exp \left\{ - \int_{t_1}^t g(t') \, \d t' \right\} g(t) \,
\frac{f(t)}{g(t)} ~,
\end{equation}
where the first exponential times $g(t_1)$ gives the probability
that $t_1$ is first selected, the square brackets the probability
that $t_1$
is subsequently rejected, the following piece the probability that
$t = t_2$ is selected when starting from $t_1$, and the final factor
that $t$ is retained. The whole is to be integrated over all
possible intermediate times $t_1$. The exponentials together give an
integral over the range from 0 to $t$, just as in ${\cal P}_0$,
and the factor for the final step being accepted is also the same,
so therefore one finds that
\begin{equation}
{\cal P}_1 (t) = {\cal P}_0 (t) \int_0^t \d t_1
\left[ g(t_1) - f(t_1) \right] ~.
\end{equation}
This generalizes.
In ${\cal P}_2$ one has to consider two intermediate times,
$0 \leq t_1 \leq t_2 \leq t_3 = t$, and so
\begin{eqnarray}
{\cal P}_2 (t) & = & {\cal P}_0 (t)
\int_0^t \d t_1 \left[ g(t_1) - f(t_1) \right]
\int_{t_1}^t \d t_2 \left[ g(t_2) - f(t_2) \right] \nonumber \\
& = & {\cal P}_0 (t) \frac{1}{2} \left(
\int_0^t \left[ g(t') - f(t') \right] \d t' \right)^2 ~.
\end{eqnarray}
The last equality is most easily seen if one also considers the
alternative region $0 \leq t_2 \leq t_1 \leq t$, where the
r\^oles of $t_1$ and $t_2$ have just been interchanged, and the
integral therefore has the same value as in the region considered.
Adding the two regions, however, the integrals over $t_1$ and
$t_2$ decouple, and become equal. In general, for ${\cal P}_i$,
the $i$ intermediate times can be ordered in $i!$ different ways.
Therefore the total probability to accept $t$, in any step, is
\begin{eqnarray}
{\cal P} (t) & = &
\displaystyle{ \sum_{i=0}^{\infty} {\cal P}_i (t)
= {\cal P}_0 (t) \sum_{i=0}^{\infty} \frac{1}{i!}
\left( \int_0^t \left[ g(t') - f(t') \right] \d t' \right)^i }
\nonumber \\
& = & \displaystyle{ f(t) \exp \left\{ - \int_0^t g(t') \, \d t'
\right\} \exp \left\{ \int_0^t \left[ g(t') - f(t') \right] \d t'
\right\} } \nonumber \\
& = & \displaystyle{ f(t) \exp \left\{ - \int_0^t f(t') \, \d t'
\right\} ~, }
\end{eqnarray}
which is the desired answer.
If the process is to be stopped at some scale $t_{\mmax}$, i.e.\
if one would like to remain with a fraction ${\cal N}(t_{\mmax})$
of events where nothing happens at all, this is easy to include
in the veto algorithm: just iterate upwards in $t$ at usual, but
stop the process if no allowed branching is found before
$t_{\mmax}$.
Usually $f(t)$ is a function also of additional variables $x$. The
methods of the preceding section are easy to generalize if one
can find a suitable function $g(t,x)$ with $f(t,x) \leq g(t,x)$.
The $g(t)$ used in the veto algorithm is the integral of $g(t,x)$
over $x$. Each time a $t_i$ has been selected also an $x_i$ is
picked, according to $g(t_i,x) \, dx$, and the $(t,x)$ point is
accepted with probability $f(t_i,x_i)/g(t_i,x_i)$.
\subsection{The Random Number Generator}
In recent years, progress has been made in constructing portable
generators with large periods and other good properties; see the review
\cite{Jam90}. Therefore the current version contains a random number
generator based on the algorithm proposed by Marsaglia, Zaman and Tsang
\cite{Mar90}. This routine should work on any machine with a mantissa
of at least 48 digits, i.e.\ on computers with a 64-bit (or more)
representation of double precision real numbers. Given the same
initial state, the sequence will also be identical on different platforms.
This need not mean that the same sequence of events will be generated,
since the different treatments of roundoff errors in numerical
operations will lead to slightly different real numbers being tested
against these random numbers in IF statements. Also code optimization
may lead to a divergence of the event sequence.
Apart from nomenclature issues, the coding of \ttt{PYR} as
a function rather than a subroutine, and the extension to double
precision, the only difference between our code and the code given in
\cite{Jam90} is that slightly different algorithms are used to ensure
that the random number is not equal to 0 or 1 within the machine
precision. Further developments of the algorithm has been proposed
\cite{Lus94} to remove residual possibilities of small long-range
correlations, at the price of a slower generation procedure. However,
given that {\Py} is using random numbers for so many different tasks,
without any fixed cycle, this has been deemed unnecessary.
The generator has a period of over $10^{43}$, and the
possibility to obtain almost $10^9$
different and disjoint subsequences, selected
by giving an initial integer number. The price to be paid for the
long period is that the state of the generator at a given moment
cannot be described by a single integer, but requires about 100 words.
Some of these are real numbers, and are thus not correctly represented
in decimal form. The old-style procedure, which made it possible to
restart the generation from a seed value written to the run output,
is therefore
not convenient. The CERN library implementation keeps track of the
number of random numbers generated since the start. With this value
saved, in a subsequent run the random generator can be asked to skip
ahead the corresponding number of random numbers. {\Py}
is a heavy user of random numbers, however: typically 30\% of the full
run time is spent on random number generation. Of this, half is
overhead coming from the function call administration, but the other
half is truly related to the speed of the algorithm. Therefore a
skipping ahead would take place with 15\% of the time cost of the
original run, i.e.\ an uncomfortably high figure.
Instead a different solution is chosen here. Two special routines are
provided for writing and reading the state of the random number
generator (plus some initialization information) on a sequential
file, in a platform-dependent internal representation. The file used
for this purpose
has to be specified by you, and opened for read and write.
A state is written as a single record, in free format. It is possible
to write an arbitrary number of states on a file, and a record can be
overwritten, if so desired. The event generation loop might then look
something like:
\begin{Enumerate}
\item save the state of the generator on file (using flag set in
point 3 below),
\item generate an event,
\item study the event for errors or other reasons why to regenerate
it later; set flag to overwrite previous generator state if no errors,
otherwise set flag to create new record;
\item loop back to point 1.
\end{Enumerate}
With this procedure, the file will contain the state before each of the
problematical events. These events can therefore be generated in a
shorter run, where further information can be printed. (Inside {\Py},
some initialization may take place in connection with the very first
event generated in a run, so it may be necessary to generate one
ordinary event before reading in a saved state to generate the
interesting events.) An alternative approach might be to save the state
every 100 events or so. If the events are subsequently processed through
a detector simulation, you may have to save also other sets of seeds,
naturally.
Unfortunately, the procedure is not always going to work. For instance,
if cross section maximum violations have occured before the interesting
event in the original run, there is a possibility that another event is
picked in the re-started one, where the maximum weight estimate has not
been updated. Another problem is the multiple interaction machinery,
where some of the options contain an element of learning, which again
means that the event sequence may be broken.
In addition to the service routines, the common block which contains
the state of the generator is available for manipulation,
if you so desire. In particular, the initial seed value is by
default 19780503, i.e.\ different from the Marsaglia/CERN default
54217137. It is possible to change this value before any random numbers
have been generated, or to force re-initialization in mid-run with any
desired new seed.
It should be noted that, of course, the appearance of a random
number generator package inside {\Py} does in no way preclude
the use of other routines. You can easily revert to having
\ttt{PYR} as nothing but an interface to an
arbitrary external random number generator; e.g.\ to call a routine
\ttt{RNDM} all you need to have is
\begin{verbatim}
FUNCTION PYR(IDUMMY)
IMPLICIT DOUBLE PRECISION(A-H, O-Z)
100 PYR=RNDM(IDUMMY)
IF(PYR.LE.0D0.OR.PYR.GE.1D0) GOTO 100
RETURN
END
\end{verbatim}
The random generator subpackage consists of the following components.
\drawbox{R = PYR(IDUMMY)}\label{p:PYR}
\begin{entry}
\itemc{Purpose:} to generate a (pseudo)random number \ttt{R} uniformly
in the range $0 <$ \ttt{R} $< 1$, i.e.\ excluding the endpoints.
\iteme{IDUMMY :} dummy input argument; normally 0.
\end{entry}
\drawbox{CALL PYRGET(LFN,MOVE)}\label{p:PYRGET}
\begin{entry}
\itemc{Purpose:} to dump the current state of the random number
generator on a separate file, using internal representation for
real and integer numbers. To be precise, the full contents of
the \ttt{PYDATR} common block are written on the file, with the
exception of \ttt{MRPY(6)}.
\iteme{LFN :} (logical file number) the file number to which the state
is dumped. You
must associate this number with a true file (with a platform-dependent
name), and see to it that this file is open for write.
\iteme{MOVE :} choice of adding a new record to the file or
overwriting old record(s). Normally only options 0 or $-$1 should be
used.
\begin{subentry}
\iteme{= 0 (or > 0) :} add a new record to the end of the
file.
\iteme{= -1 :} overwrite the last record with a new one (i.e.\ do
one \ttt{BACKSPACE} before the new write).
\iteme{= $-n$ :} back up $n$ records before writing the
new record. The records following after the new one are lost,
i.e.\ the last $n$ old records are lost and one new added.
\end{subentry}
\end{entry}
\drawbox{CALL PYRSET(LFN,MOVE)}\label{p:PYRSET}
\begin{entry}
\itemc{Purpose:} to read in a state for the random number generator,
from which the subsequent generation can proceed. The state must
previously have been saved by a \ttt{PYRGET} call. Again the full
contents of the \ttt{PYDATR} common block are read, with the
exception of \ttt{MRPY(6)}.
\iteme{LFN :} (logical file number) the file number from which the
state is read. You
must associate this number with a true file previously written with
a \ttt{PYRGET} call, and see to it that this file is open for read.
\iteme{MOVE :} positioning in file before a record is read. With zero
value, records are read one after the other for each new call, while
non-zero values may be used to navigate back and forth, and e.g.
return to the same initial state several times.
\begin{subentry}
\iteme{= 0 :} read the next record.
\iteme{= $+n$ :} skip ahead $n$ records before reading the record
that sets the state of the random number generator.
\iteme{= $-n$ :} back up $n$ records before reading the record that
sets the state of the random number generator.
\end{subentry}
\end{entry}
\drawbox{COMMON/PYDATR/MRPY(6),RRPY(100)}\label{p:PYDATR}
\begin{entry}
\itemc{Purpose:} to contain the state of the random number generator
at any moment (for communication between \ttt{PYR}, \ttt{PYRGET}
and \ttt{PYRSET}), and also to provide you with the
possibility to initialize different random number sequences, and to
know how many numbers have been generated.
\iteme{MRPY(1) :}\label{p:MRPY} (D = 19780503) the integer number that
specifies
which of the possible subsequences will be initialized in the next
\ttt{PYR} call for which \ttt{MRPY(2) = 0}. Allowed values are
0 $\leq$ \ttt{MRPY(1)} $\leq$ 900\,000\,000, the original Marsaglia
(and CERN library) seed is 54217137. The \ttt{MRPY(1)} value is not
changed by any of the {\Py} routines.
\iteme{MRPY(2) :} (D = 0) initialization flag, put to 1 in the first
\ttt{PYR} call of run. A re-initialization of the random number
generator can be made in mid-run by resetting \ttt{MRPY(2)} to 0
by hand. In addition, any time the counter \ttt{MRPY(3)} reaches
1000000000, it is reset to 0 and \ttt{MRPY(2)} is increased by 1.
\iteme{MRPY(3) :} (R) counter for the number of random numbers
generated from the beginning of the run. To avoid overflow when
very many numbers are generated, \ttt{MRPY(2)} is used as
described above.
\iteme{MRPY(4), MRPY(5) :} \ttt{I97} and \ttt{J97} of the CERN
library implementation; part of the state of the generator.
\iteme{MRPY(6) :} (R) current position, i.e.\ how many records after
beginning, in the file; used by \ttt{PYRGET} and \ttt{PYRSET}.
\iteme{RRPY(1) - RRPY(97) :}\label{p:RRPY} the \ttt{U} array of the
CERN library implementation; part of the state of the generator.
\iteme{RRPY(98) - RRPY(100) :} \ttt{C}, \ttt{CD} and \ttt{CM} of the
CERN library implementation; the first part of the state of the
generator, the latter two constants calculated at initialization.
\end{entry}
\clearpage
\section{The Event Record}
The event record is the central repository for information about
the particles produced in the current event: flavours, momenta,
event history, and production vertices. It plays a very central
r\^ole: without a proper understanding of what the record is and
how information is stored, it is meaningless to try to use {\Py}.
The record is stored in the common block
\ttt{PYJETS}. Almost all the routines that the user calls
can be viewed as performing some action on the record: fill a
new event, let partons fragment or particles decay, boost it,
list it, find clusters, etc.
In this section we will first describe the {\KF} flavour code,
subsequently the \ttt{PYJETS} common block, and then give a few
comments about the r\^ole of the event record in the programs.
To ease the interfacing of different event generators, a
\ttt{HEPEVT} standard common-block structure for the event record
has been agreed on. For historical reasons the standard common blocks
are not directly used in {\Py}, but a conversion routine comes with
the program, and is described at the end of this section.
\subsection{Particle Codes}
\label{ss:codes}
The Particle Data Group particle code \cite{PDG88,PDG92,PDG00} is
used consistently throughout the program. Almost all known
discrepancies between earlier versions of the PDG standard and the
{\Py} usage have now been resolved. The one known exception is the
(very uncertain) classification of $\f_0(980)$, with $\f_0(1370)$
also affected as a consequence. There is also a possible point of
confusion in the technicolor sector between ${\pi'}^0_{\mrm{tc}}$
and $\eta_{\mrm{tc}}$. The latter is retained for historical reasons,
whereas the former was introduced for consistency in models of
top-color-assisted technicolor.
The PDG standard, with the local {\Py} extensions, is referred to as
the {\KF} particle code. This code you have to be thoroughly familiar
with. It is described below.
The {\KF} code is not convenient for a direct storing of masses,
decay data, or other particle properties, since the {\KF}
codes are so spread out. Instead a compressed code {\KC} between
1 and 500 is used here. A particle and its antiparticle are
mapped to the same {\KC} code, but else the mapping is unique.
Normally this code is only used at very specific
places in the program, not visible to the user. If need be, the
correspondence can always be obtained by using the function
\ttt{PYCOMP}, i.e.\ \mbox{\ttt{KC = PYCOMP(KF)}}. This mapping is not
hardcoded, but can be changed by user intervention, e.g.
by introducing new particles with the \ttt{PYUPDA} facility.
It is therefore not intended that you should ever want or need to know
any {\KC} codes at all. It may be useful to know, however, that for codes
smaller than 80, {\KF} and {\KC} agree. Normally a user would never do the
inverse mapping, but we note that this is stored as
\ttt{KF = KCHG(KC,4)}, making use of the \ttt{KCHG} array in the
\ttt{PYDAT2} common block. Of course, the sign of a particle could
never be recovered by this inverse operation.
The particle names printed in the tables in this section correspond
to the ones obtained
with the routine \ttt{PYNAME}, which is used extensively, e.g.\ in
\ttt{PYLIST}. Greek characters are spelt out in full, with a capital
first letter to correspond to a capital Greek letter. Generically the
name of a particle is made up of the following pieces:
\begin{Enumerate}
\item The basic root name. This includes a * for most spin 1
($L = 0$) mesons and spin $3/2$ baryons, and a $'$ for some spin
$1/2$ baryons (where there are two states to be distinguished,
cf. $\Lambda$--$\Sigma^0$). The rules for heavy baryon naming are in
accordance with the 1986 Particle Data Group conventions \cite{PDG86}.
For mesons with one unit of orbital angular momentum, K (D, B,
\ldots) is used for quark-spin 0 and K* (D*, B*, \ldots) for
quark-spin 1 mesons; the convention for `*' may here deviate slightly
from the one used by the PDG.
\item Any lower indices, separated from the root by a \_. For
heavy hadrons, this is the additional heavy-flavour content not
inherent in the root itself. For a diquark, it is the spin.
\item The characters `bar' for an antiparticle, wherever the distinction
between particle and antiparticle is not inherent in the charge
information.
\item Charge information: $++$, $+$, $0$, $-$, or $--$.
Charge is not given for quarks or diquarks. Some neutral particles
which are customarily given without a 0 also here lack it,
such as neutrinos, $\g$, $\gamma$,
and flavour-diagonal mesons other than $\pi^0$ and $\rho^0$. Note that
charge is included both for the proton and the neutron. While
non-standard, it is helpful in avoiding misunderstandings when
looking at an event listing.
\end{Enumerate}
Below follows a list of {\KF} particle codes. The list is not complete;
a more extensive one may be obtained with \ttt{CALL PYLIST(11)}.
Particles are grouped together, and the basic rules are described
for each group. Whenever a distinct antiparticle exists, it is given
the same {\KF} code with a minus sign (whereas {\KC} codes are always
positive).
\begin{Enumerate}
\item Quarks and leptons, Table \ref{t:codeone}. \\
This group contains the basic building blocks of matter, arranged
according to family, with the lower member of weak isodoublets also
having the smaller code (thus $\d$ precedes $\u$). A fourth generation
is included as part of the scenarios for exotic physics. The quark codes
are used as building blocks for the diquark, meson and baryon codes below.
\begin{table}[ptb]
\caption{Quark and lepton codes.
\protect\label{t:codeone} }
\begin{center}
\begin{tabular}{|c|c|c||c|c|c|}
\hline
{\KF} & Name & Printed & {\KF} & Name & Printed \\
\hline
1 & $\d$ & \ttt{d} & 11 & $\e^-$ & \ttt{e-} \\
2 & $\u$ & \ttt{u} & 12 & $\nu_{\e}$ & \ttt{nu\_e} \\
3 & $\s$ & \ttt{s} & 13 & $\mu^-$ & \ttt{mu-} \\
4 & $\c$ & \ttt{c} & 14 & $\nu_{\mu}$ & \ttt{nu\_mu} \\
5 & $\b$ & \ttt{b} & 15 & $\tau^-$ & \ttt{tau-} \\
6 & $\t$ & \ttt{t} & 16 & $\nu_{\tau}$ & \ttt{nu\_tau} \\
7 & $\b'$ & \ttt{b'} & 17 & $\tau'$ & \ttt{tau'} \\
8 & $\t'$ & \ttt{t'} & 18 & $\nu'_{\tau}$ & \ttt{nu'\_tau} \\
9 & & & 19 & & \\
10 & & & 20 & & \\
\hline
\end{tabular}
\end{center}
\end{table}
\item Gauge bosons and other fundamental bosons,
Table \ref{t:codetwo}. \\
This group includes all the gauge and Higgs bosons of the
Standard Model, as well as some of the bosons appearing in
various extensions of it. They correspond to one extra {\bf U(1)}
and one extra {\bf SU(2)} group, a further Higgs doublet,
a graviton, a horizontal gauge boson $\R$ (coupling between
families), and a (scalar) leptoquark $\L_{\Q}$.
\begin{table}[ptb]
\caption{Gauge boson and other fundamental boson codes.
\protect\label{t:codetwo} }
\begin{center}
\begin{tabular}{|c|c|c||c|c|c|}
\hline
{\KF} & Name & Printed & {\KF} & Name & Printed \\
\hline
21 & $\g$ & \ttt{g} & 31 & & \\
22 & $\gamma$ & \ttt{gamma} & 32 & $\Z'^0$ & \ttt{Z'0} \\
23 & $\Z^0$ & \ttt{Z0} & 33 & $\Z''^0$ & \ttt{Z"0}\\
24 & $\W^+$ & \ttt{W+} & 34 & $\W'^+$ & \ttt{W'+} \\
25 & $\hrm^0$ & \ttt{h0} & 35 & $\H^0$ & \ttt{H0} \\
26 & & & 36 & $\A^0$ & \ttt{A0} \\
27 & & & 37 & $\H^+$ & \ttt{H+} \\
28 & & & 38 & & \\
29 & & & 39 & $\mrm{G}$ & \ttt{Graviton} \\
30 & & & 40 & & \\
& & & 41 & $\R^0$ & \ttt{R0} \\
& & & 42 & $\L_{\Q}$ & \ttt{LQ}\\
\hline
\end{tabular}
\end{center}
\end{table}
\item Exotic particle codes. \\
The positions 43--80 are used as temporary sites for exotic particles
that eventually may be shifted to a separate code sequence. Currently
this list only consists of the particle codes 45 and 46, described
among the Supersymmetric codes below. The ones not in use are at your
disposal, but with no guarantees that they will remain so.
\item Various special codes, Table \ref{t:codefour}. \\
In a Monte Carlo, it is always necessary to have codes that do not
correspond to any specific particle, but are used to lump together
groups of similar particles for decay treatment (nowadays largely
obsolete), to specify generic decay products (also obsolete),
or generic intermediate states in external processes, or
additional event record information from jet searches. These codes,
which again are non-standard, are found between numbers 81 and 100.\\
The junction, code 88, is not a physical particle but marks the place
in the event record where three string pieces come together in a point,
e.g. a Y-shaped topology with a quark at each end. No distinction is made
between a junction and an antijunction, i.e.\ whether a baryon or an
antibaryon is going to be produced in the neighbourhood of the junction.
\begin{table}[ptb]
\caption{Various special codes.
\protect\label{t:codefour} }
\begin{center}
\begin{tabular}{|c|c|c|}
\hline
{\KF} & Printed & Meaning \\
\hline
81 & \ttt{specflav} & Spectator flavour; used in decay-product
listings \\
82 & \ttt{rndmflav} & A random $\u$, $\d$, or $\s$ flavour;
possible decay product \\
83 & \ttt{phasespa} & Simple isotropic phase-space decay \\
84 & \ttt{c-hadron} & Information on decay of generic charm
hadron \\
85 & \ttt{b-hadron} & Information on decay of generic bottom
hadron \\
86 & & \\
87 & & \\
88 & \ttt{junction} & A junction of three string pieces \\
89 & & (internal use for unspecified resonance data) \\
90 & \ttt{system} & Intermediate pseudoparticle in external process \\
91 & \ttt{cluster} & Parton system in cluster fragmentation \\
92 & \ttt{string} & Parton system in string fragmentation \\
93 & \ttt{indep.} & Parton system in independent fragmentation \\
94 & \ttt{CMshower} & Four-momentum of time-like showering system \\
95 & \ttt{SPHEaxis} & Event axis found with \ttt{PYSPHE} \\
96 & \ttt{THRUaxis} & Event axis found with \ttt{PYTHRU} \\
97 & \ttt{CLUSjet} & Jet (cluster) found with \ttt{PYCLUS} \\
98 & \ttt{CELLjet} & Jet (cluster) found with \ttt{PYCELL} \\
99 & \ttt{table} & Tabular output from \ttt{PYTABU} \\
100 & & \\
\hline
\end{tabular}
\end{center}
\end{table}
\item Diquark codes, Table \ref{t:codefive}. \\
A diquark made up of a quark with code $i$ and another with code $j$,
where $i \geq j$, and with total spin $s$, is given the code
\begin{equation}
\mtt{KF} = 1000 i + 100 j + 2s + 1 ~,
\end{equation}
i.e.\ the tens position is left empty (cf. the baryon code below).
Some of the most frequently used codes are listed in the table. All
the lowest-lying spin 0 and 1 diquarks are included in the program.
\begin{table}[ptb]
\caption{Diquark codes. For brevity, diquarks containing $\c$ or $\b$
quarks are not listed, but are defined analogously.
\protect\label{t:codefive} }
\begin{center}
\begin{tabular}{|c|c|c||c|c|c|}
\hline
{\KF} & Name & Printed & {\KF} & Name & Printed \\
\hline
& & & 1103 & $\d\d_1$ & \ttt{dd\_1} \\
2101 & $\u\d_0$ & \ttt{ud\_0} & 2103 & $\u\d_1$ & \ttt{ud\_1} \\
& & & 2203 & $\u\u_1$ & \ttt{uu\_1} \\
3101 & $\s\d_0$ & \ttt{sd\_0} & 3103 & $\s\d_1$ & \ttt{sd\_1} \\
3201 & $\s\u_0$ & \ttt{su\_0} & 3203 & $\s\u_1$ & \ttt{su\_1} \\
& & & 3303 & $\s\s_1$ & \ttt{ss\_1} \\
\hline
\end{tabular}
\end{center}
\end{table}
\item Meson codes, Tables \ref{t:codesixa} and \ref{t:codesixb}. \\
A meson made up of a quark with code $i$ and an antiquark with
code $-j$, $j \neq i$, and with total spin $s$, is given the code
\begin{equation}
\mtt{KF} = \left\{ 100 \max(i,j) + 10 \min(i,j) + 2s + 1 \right\}
\, \mrm{sign}(i-j) \, (-1)^{\max(i,j)} ~,
\label{eq:KFmeson}
\end{equation}
assuming it is not orbitally or radially excited.
Note the presence of an extra $-$ sign if the heaviest quark is a
down-type one. This is in accordance with the particle--antiparticle
distinction adopted in the 1986 Review of Particle Properties
\cite{PDG86}. It means for example that a $\B$ meson contains a
$\bbar$ antiquark rather than a $\b$ quark.
The flavour-diagonal states are arranged in order of ascending
mass. Thus the obvious generalization of eq.~(\ref{eq:KFmeson})
to $\mtt{KF} = 110 i + 2 s + 1$ is only valid for charm and bottom.
The lighter quark states can appear mixed, e.g. the $\pi^0$
(111) is an equal mixture of $\d\dbar$ (na\"{\i}vely code 111) and
$\u\ubar$ (na\"{\i}vely code 221).
The standard rule of having the last digit of the form
$2s+1$ is broken for the $\K_{\mrm{S}}^0$--$\K_{\mrm{L}}^0$ system,
where it is 0, and this convention should carry over to mixed states
in the $\B$ meson system, should one choose to define such. For
higher multiplets with the same spin, $\pm$10000, $\pm$20000, etc.,
are added to provide the extra distinction needed. Some of the most
frequently used codes are given below.
The full lowest-lying pseudoscalar and vector multiplets are included
in the program, Table \ref{t:codesixa}.
\begin{table}[ptb]
\caption{Meson codes, part 1.
\protect\label{t:codesixa} }
\begin{center}
\begin{tabular}{|c|c|c||c|c|c|}
\hline
{\KF} & Name & Printed & {\KF} & Name & Printed \\
\hline
211 & $\pi^+$ & \ttt{pi+} & 213 & $\rho^+$ & \ttt{rho+} \\
311 & $\K^0$ & \ttt{K0} & 313 & $\K^{*0}$ & \ttt{K*0} \\
321 & $\K^+$ & \ttt{K+} & 323 & $\K^{*+}$ & \ttt{K*+} \\
411 & $\D^+$ & \ttt{D+} & 413 & $\D^{*+}$ & \ttt{D*+} \\
421 & $\D^0$ & \ttt{D0} & 423 & $\D^{*0}$ & \ttt{D*0} \\
431 & $\D_{\s}^+$ & \ttt{D\_s+} &
433 & $\D_{\s}^{*+}$ & \ttt{D*\_s+} \\
511 & $\B^0$ & \ttt{B0} & 513 & $\B^{*0}$ & \ttt{B*0} \\
521 & $\B^+$ & \ttt{B+} & 523 & $\B^{*+}$ & \ttt{B*+} \\
531 & $\B_{\s}^0$ & \ttt{B\_s0} &
533 & $\B_{\s}^{*0}$ & \ttt{B*\_s0} \\
541 & $\B_{\c}^+$ & \ttt{B\_c+} &
543 & $\B_{\c}^{*+}$ & \ttt{B*\_c+} \\
111 & $\pi^0$ & \ttt{pi0} & 113 & $\rho^0$ & \ttt{rho0} \\
221 & $\eta$ & \ttt{eta} & 223 & $\omega$ & \ttt{omega} \\
331 & $\eta'$ & \ttt{eta'} & 333 & $\phi$ & \ttt{phi} \\
441 & $\eta_{\c}$ & \ttt{eta\_c} & 443 & $\Jpsi$ & \ttt{J/psi} \\
551 & $\eta_{\b}$ & \ttt{eta\_b} &
553 & $\Upsilon$ & \ttt{Upsilon} \\
\hline
130 & $\K_{\mrm{L}}^0$ & \ttt{K\_L0} & & & \\
310 & $\K_{\mrm{S}}^0$ & \ttt{K\_S0} & & & \\
\hline
\end{tabular}
\end{center}
\end{table}
Also the lowest-lying orbital angular momentum $L = 1$ mesons are
included, Table \ref{t:codesixb}: one pseudovector multiplet
obtained for total quark-spin 0 ($L = 1, S = 0 \Rightarrow J = 1$)
and one scalar, one pseudovector and one tensor multiplet obtained
for total quark-spin 1 ($L = 1, S = 1 \Rightarrow J = 0, 1$ or 2),
where $J$ is what is conventionally called the spin $s$ of the meson.
Any mixing between the two pseudovector multiplets is
not taken into account. Please note that some members of these
multiplets have still not been found, and are included here only based
on guesswork. Even for known ones, the information on particles
(mass, width, decay modes) is highly incomplete.
Only two radial excitations are included, the $\psi' = \psi(2S)$ and
$\Upsilon' = \Upsilon(2S)$.
\begin{table}[ptb]
\caption{Meson codes, part 2. For brevity, states with $\b$ quark
are omitted from this listing, but are defined in the program.
\protect\label{t:codesixb} }
\begin{center}
\begin{tabular}{|c|c|c||c|c|c|}
\hline
{\KF} & Name & Printed & {\KF} & Name & Printed \\
\hline
10213 & $\b_1$ & \ttt{b\_1+} &
10211 & $\a_0^+$ & \ttt{a\_0+} \\
10313 & $\K_1^0$ & \ttt{K\_10} &
10311 & $\K_0^{*0}$ & \ttt{K*\_00} \\
10323 & $\K_1^+$ & \ttt{K\_1+} &
10321 & $\K_0^{*+}$ & \ttt{K*\_0+} \\
10413 & $\D_1^+$ & \ttt{D\_1+} &
10411 & $\D_0^{*+}$ & \ttt{D*\_0+} \\
10423 & $\D_1^0$ & \ttt{D\_10} &
10421 & $\D_0^{*0}$ & \ttt{D*\_00} \\
10433 & $\D_{1 \s}^+$ & \ttt{D\_1s+} &
10431 & $\D_{0 \s}^{*+}$ & \ttt{D*\_0s+} \\
10113 & $\b_1^0$ & \ttt{b\_10} &
10111 & $\a_0^0$ & \ttt{a\_00} \\
10223 & $\hrm_1^0$ & \ttt{h\_10} &
10221 & $\f_0^0$ & \ttt{f\_00} \\
10333 & $\hrm'^0_1$ & \ttt{h'\_10} &
10331 & $\f'^0_0$ & \ttt{f'\_00} \\
10443 & $\hrm_{1 \c}^0$ & \ttt{h\_1c0} &
10441 & $\chi_{0 \c}^0$ & \ttt{chi\_0c0} \\ \hline
20213 & $\a_1^+$ & \ttt{a\_1+} &
215 & $\a_2^+$ & \ttt{a\_2+} \\
20313 & $\K_1^{*0}$ & \ttt{K*\_10} &
315 & $\K_2^{*0}$ & \ttt{K*\_20} \\
20323 & $\K_1^{*+}$ & \ttt{K*\_1+} &
325 & $\K_2^{*+}$ & \ttt{K*\_2+} \\
20413 & $\D_1^{*+}$ & \ttt{D*\_1+} &
415 & $\D_2^{*+}$ & \ttt{D*\_2+} \\
20423 & $\D_1^{*0}$ & \ttt{D*\_10} &
425 & $\D_2^{*0}$ & \ttt{D*\_20} \\
20433 & $\D_{1 \s}^{*+}$ & \ttt{D*\_1s+} &
435 & $\D_{2 \s}^{*+}$ & \ttt{D*\_2s+} \\
20113 & $\a_1^0$ & \ttt{a\_10} &
115 & $\a_2^0$ & \ttt{a\_20} \\
20223 & $\f_1^0$ & \ttt{f\_10} &
225 & $\f_2^0$ & \ttt{f\_20} \\
20333 & $\f'^0_1$ & \ttt{f'\_10} &
335 & $\f'^0_2$ & \ttt{f'\_20} \\
20443 & $\chi_{1 \c}^0$ & \ttt{chi\_1c0} &
445 & $\chi_{2 \c}^0$ & \ttt{chi\_2c0} \\ \hline
100443 & $\psi'$ & \ttt{psi'} & & & \\
100553 & $\Upsilon'$ & \ttt{Upsilon'} & & & \\
\hline
\end{tabular}
\end{center}
\end{table}
\item Baryon codes, Table \ref{t:codeseven}. \\
A baryon made up of quarks $i$, $j$ and $k$, with $i \geq j \geq k$,
and total spin $s$, is given the code
\begin{equation}
\mtt{KF} = 1000 i + 100 j + 10 k + 2s + 1 ~.
\end{equation}
An exception is provided by spin $1/2$ baryons made up of three
different types of quarks, where the two lightest quarks form a spin-0
diquark ($\Lambda$-like baryons). Here the order of the $j$ and $k$
quarks is reversed, so as to provide a simple means of distinction
to baryons with the lightest quarks in a spin-1 diquark
($\Sigma$-like baryons).
For hadrons with heavy flavours, the root names are Lambda or
Sigma for hadrons with two $\u$ or $\d$ quarks, Xi for those
with one, and Omega for those without $\u$ or $\d$ quarks.
Some of the most frequently used codes are given in Table
\ref{t:codeseven}. The full lowest-lying spin $1/2$ and $3/2$
multiplets are included in the program.
\begin{table}[ptb]
\caption{Baryon codes. For brevity, some states with $\b$ quarks
or multiple $\c$ ones are omitted from this listing, but are
defined in the program.
\protect\label{t:codeseven} }
\begin{center}
\begin{tabular}{|c|c|c||c|c|c|}
\hline
{\KF} & Name & Printed & {\KF} & Name & Printed \\
\hline
& & &
1114 & $\Delta^-$ & \ttt{Delta-} \\
2112 & $\n$ & \ttt{n0} &
2114 & $\Delta^0$ & \ttt{Delta0} \\
2212 & $\p$ & \ttt{p+} &
2214 & $\Delta^+$ & \ttt{Delta+} \\
& & &
2224 & $\Delta^{++}$ & \ttt{Delta++} \\
3112 & $\Sigma^-$ & \ttt{Sigma-} &
3114 & $\Sigma^{*-}$ & \ttt{Sigma*-} \\
3122 & $\Lambda^0$ & \ttt{Lambda0} & & & \\
3212 & $\Sigma^0$ & \ttt{Sigma0} &
3214 & $\Sigma^{*0}$ & \ttt{Sigma*0} \\
3222 & $\Sigma^+$ & \ttt{Sigma+} &
3224 & $\Sigma^{*+}$ & \ttt{Sigma*+} \\
3312 & $\Xi^-$ & \ttt{Xi-} &
3314 & $\Xi^{*-}$ & \ttt{Xi*-} \\
3322 & $\Xi^0$ & \ttt{Xi0} &
3324 & $\Xi^{*0}$ & \ttt{Xi*0} \\
& & &
3334 & $\Omega^-$ & \ttt{Omega-} \\
4112 & $\Sigma_{\c}^0$ & \ttt{Sigma\_c0} &
4114 & $\Sigma_{\c}^{*0}$ & \ttt{Sigma*\_c0} \\
4122 & $\Lambda_{\c}^+$ & \ttt{Lambda\_c+} & & & \\
4212 & $\Sigma_{\c}^+$ & \ttt{Sigma\_c+} &
4214 & $\Sigma_{\c}^{*+}$ & \ttt{Sigma*\_c+} \\
4222 & $\Sigma_{\c}^{++}$ & \ttt{Sigma\_c++} &
4224 & $\Sigma_{\c}^{*++}$ & \ttt{Sigma*\_c++} \\
4132 & $\Xi_{\c}^0$ & \ttt{Xi\_c0} & & & \\
4312 & $\Xi'^0_{\c}$ & \ttt{Xi'\_c0} &
4314 & $\Xi_{\c}^{*0}$ & \ttt{Xi*\_c0} \\
4232 & $\Xi_{\c}^+$ & \ttt{Xi\_c+} & & & \\
4322 & $\Xi'^+_{\c}$ & \ttt{Xi'\_c+} &
4324 & $\Xi_{\c}^{*+}$ & \ttt{Xi*\_c+} \\
4332 & $\Omega_{\c}^0$ & \ttt{Omega\_c0} &
4334 & $\Omega_{\c}^{*0}$ & \ttt{Omega*\_c0} \\
5112 & $\Sigma_{\b}^-$ & \ttt{Sigma\_b-} &
5114 & $\Sigma_{\b}^{*-}$ & \ttt{Sigma*\_b-} \\
5122 & $\Lambda_{\b}^0$ & \ttt{Lambda\_b0} & & & \\
5212 & $\Sigma_{\b}^0$ &\ttt{Sigma\_b0} &
5214 & $\Sigma_{\b}^{*0}$ & \ttt{Sigma*\_b0} \\
5222 & $\Sigma_{\b}^+$ & \ttt{Sigma\_b+} &
5224 & $\Sigma_{\b}^{*+}$ & \ttt{Sigma*\_b+} \\
\hline
\end{tabular}
\end{center}
\end{table}
\item QCD effective states, Table \ref{t:codeeight}. \\
We here include the pomeron $\pomeron$ and reggeon $\reggeon$
`particles', which are important e.g.\ in the description of
diffractive scattering, but do not have a simple
correspondence with other particles in the classification scheme.\\
Also included are codes to be used for denoting diffractive states
in {\Py}, as part of the event history. The first two digits here
are 99 to denote the non-standard character. The second, third and
fourth last digits give flavour content, while the very last one is
0, to denote the somewhat unusual character of the code. Only a few
codes have been introduced with names; depending on circumstances
these also have to double up for other diffractive states. Other
diffractive codes for strange mesons and baryon beams are also
accepted by the program, but do not give nice printouts.
\begin{table}[ptb]
\caption{QCD effective states.
\protect\label{t:codeeight} }
\begin{center}
\begin{tabular}{|c|c|c|}
\hline
{\KF} & Printed & Meaning \\
\hline
110 & \ttt{reggeon} & reggeon $\reggeon$ \\
990 & \ttt{pomeron} & pomeron $\pomeron$ \\
9900110 & \ttt{rho\_diff0} & Diffractive
$\pi^0 / \rho^0 / \gamma$ state \\
9900210 & \ttt{pi\_diffr+} & Diffractive $\pi^+$ state \\
9900220 & \ttt{omega\_di0} & Diffractive $\omega$ state \\
9900330 & \ttt{phi\_diff0} & Diffractive $\phi$ state \\
9900440 & \ttt{J/psi\_di0} & Diffractive $\Jpsi$ state \\
9902110 & \ttt{n\_diffr} & Diffractive $\n$ state \\
9902210 & \ttt{p\_diffr+} & Diffractive $\p$ state \\
\hline
\end{tabular}
\end{center}
\end{table}
\item Supersymmetric codes, Table \ref{t:codenine}. \\
SUSY doubles the number of states of the Standard Model (at least).
Fermions have separate superpartners to the left- and right-handed
components. In the third generation these
are assumed to mix to nontrivial mass eigenstates, while mixing is not
included in the first two. Note that all sparticle names begin with a tilde.
Default masses are arbitrary and branching ratios not set at all.
This is taken care of at initialization if \ttt{IMSS(1)} is positive.
The $\H^0_3$, $\A^0_2$ and $\chio^0_5$ states at the bottom of the table
only appear in the Next-to-Minimal Supersymmetric Standard Model (NMSSM),
do not have standardized codes and are not fully implemented in
{\Py}, but can optionally (see \ttt{IMSS(13)}) be used in the context
of interfaces to other programs.
\begin{table}[ptb]
\caption{Supersymmetric codes.
\protect\label{t:codenine} }
\begin{center}
\begin{tabular}{|c|c|c||c|c|c|}
\hline
{\KF} & Name & Printed & {\KF} & Name & Printed \\
\hline
1000001 & $\sd_L$ & $\sim$\ttt{d\_L} &
2000001 & $\sd_R$ & $\sim$\ttt{d\_R} \\
1000002 & $\su_L$ & $\sim$\ttt{u\_L} &
2000002 & $\su_R$ & $\sim$\ttt{u\_R} \\
1000003 & $\sst_L$ & $\sim$\ttt{s\_L} &
2000003 & $\sst_R$ & $\sim$\ttt{s\_R} \\
1000004 & $\sch_L$ & $\sim$\ttt{c\_L} &
2000004 & $\sch_R$ & $\sim$\ttt{c\_R} \\
1000005 & $\sbo_1$ & $\sim$\ttt{b\_1} &
2000005 & $\sbo_2$ & $\sim$\ttt{b\_2} \\
1000006 & $\st_1$ & $\sim$\ttt{t\_1} &
2000006 & $\st_2$ & $\sim$\ttt{t\_2} \\
1000011 & $\se_L$ & $\sim$\ttt{e\_L-} &
2000011 & $\se_R$ & $\sim$\ttt{e\_R-} \\
1000012 & $\snu_{{\e}L}$ & $\sim$\ttt{nu\_eL} &
2000012 & $\snu_{{\e}R}$ & $\sim$\ttt{nu\_eR} \\
1000013 & $\smu_L$ & $\sim$\ttt{mu\_L-} &
2000013 & $\smu_R$ & $\sim$\ttt{mu\_R-} \\
1000014 & $\snu_{{\mu}L}$ & $\sim$\ttt{nu\_muL} &
2000014 & $\snu_{{\mu}R}$ & $\sim$\ttt{nu\_muR} \\
1000015 & $\stau_1$ & $\sim$\ttt{tau\_L-} &
2000015 & $\stau_2$ & $\sim$\ttt{tau\_R-} \\
1000016 & $\snu_{{\tau}L}$ & $\sim$\ttt{nu\_tauL} &
2000016 & $\snu_{{\tau}R}$ & $\sim$\ttt{nu\_tauR} \\
\hline
1000021 & $\glu$ & $\sim$\ttt{g} &
1000025 & $\chio^0_3$ & $\sim$\ttt{chi\_30} \\
1000022 & $\chio^0_1$ & $\sim$\ttt{chi\_10} &
1000035 & $\chio^0_4$ & $\sim$\ttt{chi\_40}\\
1000023 & $\chio^0_2$ & $\sim$\ttt{chi\_20} &
1000037 & $\chio^+_2$ & $\sim$\ttt{chi\_2+} \\
1000024 & $\chio^+_1$ & $\sim$\ttt{chi\_1+} &
1000039 & $\grav$ & $\sim$\ttt{Gravitino} \\
\hline
45 & $\H^0_3$ & \ttt{H\_30} &
1000045 & $\chio^0_5$ & $\sim$\ttt{chi\_50} \\
46 & $\A^0_2$ & \ttt{A\_20} & & & \\
\hline
\end{tabular}
\end{center}
\end{table}
\item Technicolor codes, Table \ref{t:codeten}. \\
A set of colourless and coloured technihadrons have been
included. The colourless technivector mesons and most
of the colourless technipions are associated with the
original strawman model of technicolor.
The coloured technirho mesons (\KF$=3100113, 3200113, 3300113
$ and $3100113$), a Coloron (or $\mathrm{V}_8$),
and additional colour singlet (\KF$=3100111$) and
colour octet (\KF=$3100111$) technipions arise
in the extended model of Topcolor assisted Technicolor (TC2).
Additional indices on these technihadrons refer
to two strongly interacting groups
\textbf{SU(3)}$_1 \times $\textbf{SU(3)}$_2$, one
for the first two generations and a second for the third generation,
which is broken down to ordinary
\textbf{SU(3)}$_\mathrm{C}$.\\
The $\eta_{\mrm{tc}}$ belongs to an older iteration of
Technicolor modelling than the rest. It was originally given the
3000221 code, and thereby now comes to clash with the
${\pi'}^0_{\mrm{tc}}$ of the current main scenario. Since the
$\eta_{\mrm{tc}}$ is one-of-a-kind, it was deemed better to move
it to make way for the ${\pi'}^0_{\mrm{tc}}$. This
leads to a slight inconsistency with the PDG codes.
\begin{table}[ptb]
\caption{Technicolor codes.
\protect\label{t:codeten} }
\begin{center}
\begin{tabular}{|c|c|c||c|c|c|}
\hline
{\KF} & Name & Printed & {\KF} & Name & Printed \\
\hline
3000111 & $\pi^0_{\mrm{tc}}$ & \ttt{pi\_tc0} &
3100021 & $\mathrm{V}_{8,\mrm{tc}}$ & \ttt{V8\_tc} \\
3000211 & $\pi^+_{\mrm{tc}}$ & \ttt{pi\_tc+} &
3100111 & $\pi^0_{22,1,\mrm{tc}}$ & \ttt{pi\_22\_1\_tc} \\
3000221 & ${\pi'}^0_{\mrm{tc}}$ & \ttt{pi'\_tc0} &
3200111 & $\pi^0_{22,8,\mrm{tc}}$ & \ttt{pi\_22\_8\_tc} \\
3000113 & $\rho^0_{\mrm{tc}}$ & \ttt{rho\_tc0} &
3100113 & $\rho^0_{11,\mrm{tc}}$ & \ttt{rho\_11\_tc} \\
3000213 & $\rho^+_{\mrm{tc}}$ & \ttt{rho\_tc+} &
3200113 & $\rho^0_{12,\mrm{tc}}$ & \ttt{rho\_12\_tc} \\
3000223 & $\omega^0_{\mrm{tc}}$ & \ttt{omega\_tc0} &
3300113 & $\rho^0_{21,\mrm{tc}}$ & \ttt{rho\_21\_tc} \\
3000331 & $\eta_{\mrm{tc}}$ & \ttt{eta\_tc0} &
3400113 & $\rho^0_{22,\mrm{tc}}$ & \ttt{rho\_22\_tc} \\
\hline
\end{tabular}
\end{center}
\end{table}
\item Excited fermion codes, Table \ref{t:codeeleven}. \\
A first generation of excited fermions are included.
\begin{table}[ptb]
\caption{Excited fermion codes.
\protect\label{t:codeeleven} }
\begin{center}
\begin{tabular}{|c|c|c||c|c|c|}
\hline
{\KF} & Name & Printed & {\KF} & Name & Printed \\
\hline
4000001 & $\u^*$ & \ttt{d*} &
4000011 & $\e^*$ & \ttt{e*-} \\
4000002 & $\d^*$ & \ttt{u*} &
4000012 & $\nu^*_{\e}$ & \ttt{nu*\_e0} \\
\hline
\end{tabular}
\end{center}
\end{table}
\item Exotic particle codes, Table \ref{t:codetwelve}. \\
This section includes the excited graviton, as the first
(but probably not last) manifestation of the possibility of
large extra dimensions. Although it is not yet in the PDG
standard, we assume that such states will go in a new series
of numbers.\\
Included is also a set of particles associated with an extra
\tbf{SU(2)} gauge group for right-handed states, as required
in order to obtain a left--right symmetric theory at high
energies. This includes right-handed (Majorana) neutrinos,
right-handed $\Z_R^0$ and $\W_R^{\pm}$ gauge bosons, and both
left- and right-handed doubly charged Higgs bosons.
Such a scenario would also contain other Higgs states, but
these do not bring anything new relative to the ones
already introduced, from an observational point of view.
Here the first two digits are 99 to denote the non-standard
character.
\begin{table}[ptb]
\caption{Exotic particle codes.
\protect\label{t:codetwelve} }
\begin{center}
\begin{tabular}{|c|c|c||c|c|c|}
\hline
{\KF} & Name & Printed & {\KF} & Name & Printed \\
\hline
5000039 & $\G^*$ & \ttt{Graviton*} & & & \\
\hline
9900012 & $\nu_{R\mrm{e}}$ & \ttt{nu\_Re} &
9900023 & $\Z_R^0$ & \ttt{Z\_R0} \\
9900014 & $\nu_{R\mu}$ & \ttt{nu\_Rmu} &
9900024 & $\W_R^+$ & \ttt{W\_R+} \\
9900016 & $\nu_{R\tau}$ & \ttt{nu\_Rtau} &
9900041 & $\H_L^{++}$ & \ttt{H\_L++} \\
& & &
9900042 & $\H_R^{++}$ & \ttt{H\_R++} \\
\hline
\end{tabular}
\end{center}
\end{table}
\item Colour octet state codes, Table \ref{t:codethirteen}. \\
Within the colour octet approach to charmonium and bottomonium
production, intermediate colour octet states can be produced and
subsequently `decay' to the normal singlet states by soft-gluon
emission. The codes have been chosen 9900000 bigger than the respective
colour-singlet state, so that they occur among the generator-specific
codes. The names are based on spectroscopic notation, with additional
upper index $(8)$ to reflect the colour octet nature.
\begin{table}[ptb]
\caption{Colour octet state codes.
\protect\label{t:codethirteen} }
\begin{center}
\begin{tabular}{|c|c|c||c|c|c|}
\hline
{\KF} & Name & Printed & {\KF} & Name & Printed \\
\hline
9900443 & $\c\cbar[^3S_1^{(8)}]$ & \ttt{cc}$\sim$\ttt{[3S18]} &
9900553 & $\b\bbar[^3S_1^{(8)}]$ & \ttt{bb}$\sim$\ttt{[3S18]} \\
9900441 & $\c\cbar[^1S_0^{(8)}]$ & \ttt{cc}$\sim$\ttt{[1S08]} &
9900551 & $\b\bbar[^1S_0^{(8)}]$ & \ttt{bb}$\sim$\ttt{[1S08]} \\
9910443 & $\c\cbar[^3P_0^{(8)}]$ & \ttt{cc}$\sim$\ttt{[3P08]} &
9910553 & $\b\bbar[^3P_0^{(8)}]$ & \ttt{bb}$\sim$\ttt{[3P08]} \\
\hline
\end{tabular}
\end{center}
\end{table}
\end{Enumerate}
A hint on large particle numbers: if you want to avoid mistyping
the number of zeros, it may pay off to define a statement like
\vspace{-0.5\baselineskip}
\begin{verbatim}
PARAMETER (KSUSY1=1000000,KSUSY2=2000000,KTECHN=3000000,
&KEXCIT=4000000,KDIMEN=5000000)
\end{verbatim}
\vspace{-0.5\baselineskip}
at the beginning of your program and then refer to particles as
\ttt{KSUSY1+1} = $\sd_L$ and so on. This then also agrees with the
internal notation (where feasible).
\subsection{The Event Record}
\label{ss:evrec}
Each new event generated is in its entirety stored in the common block
\ttt{PYJETS}, which thus forms the event record. Here each parton or
particle that appears at some stage of the fragmentation or decay
chain will occupy one line in the matrices. The different components
of this line will tell which parton/particle it is, from where it
originates, its present status (fragmented/decayed or not), its
momentum, energy and mass, and the space--time position of its
production vertex. Note that \ttt{K(I,3)}--\ttt{K(I,5)} and the
\ttt{P} and \ttt{V} vectors may take special meaning for some
specific applications (e.g.\ sphericity or cluster
analysis), as described in those connections.
The event history information stored in \ttt{K(I,3)}--\ttt{K(I,5)}
should not be taken too literally. In the particle decay chains, the
meaning of a mother is well-defined, but the fragmentation description
is more complicated. The primary hadrons produced in string
fragmentation come from the string as a whole, rather than from an
individual parton. Even when the string is not included in the history
(see \ttt{MSTU(16)}), the pointer from hadron to parton is deceptive.
For instance, in a $\q\g\qbar$ event, those hadrons are pointing towards
the $\q$ ($\qbar$) parton that were produced by fragmentation from that
end of the string, according to the random procedure used in the
fragmentation routine. No particles point to the $\g$. This assignment
seldom agrees with the visual impression, and is not intended to.
The common block \ttt{PYJETS} has expanded with time, and can now house
4000 entries. This figure may seem ridiculously large, but actually the
previous limit of 2000 was often reached in studies of high-$\pT$
processes at the LHC (and SSC). This is because the event record
contains not only the final particles, but also all intermediate partons
and hadrons, which subsequently showered, fragmented or decayed. Included
are also a wealth of photons coming from $\pi^0$ decays; the simplest
way of reducing the size of the event record is actually to switch off
$\pi^0$ decays by \ttt{MDCY(PYCOMP(111),1) = 0}. Also note that some
routines, such as \ttt{PYCLUS} and \ttt{PYCELL}, use memory after the
event record proper as a working area. Still, to change the size of
the common block, upwards or downwards, is easy: just do a global
substitute in the common block and change the \ttt{MSTU(4)} value to the
new number. If more than 10000 lines are to be used, the packing of
colour information should also be changed, see \ttt{MSTU(5)}.
\drawbox{COMMON/PYJETS/N,NPAD,K(4000,5),P(4000,5),V(4000,5)}\label{p:PYJETS}
\begin{entry}
\itemc{Purpose:} to contain the event record, i.e.\ the complete list
of all partons and particles (initial, intermediate and final) in the
current event. (By parton we here mean the subclass of particles that
carry colour, for which extra colour flow information is then required.
Normally this means quarks and gluons, which can fragment to hadrons,
but also squarks and other exotic particles fall in this category.)
\iteme{N :}\label{p:N} number of lines in the \ttt{K}, \ttt{P} and
\ttt{V} matrices occupied by the current event. \ttt{N} is continuously
updated as the definition of the original configuration and the
treatment of fragmentation and decay proceed. In the following,
the individual parton/particle number, running between 1 and
\ttt{N}, is called \ttt{I}.
\iteme{NPAD :} dummy to ensure an even number of integers before the
double precision reals, as required by some compilers.
\iteme{K(I,1) :}\label{p:K} status code KS, which gives the current
status of the parton/particle stored in the line. The ground rule is
that codes 1--10 correspond to currently existing partons/particles,
while larger codes contain partons/particles which no longer exist,
or other kinds of event information.
\begin{subentry}
\iteme{= 0 :} empty line.
\iteme{= 1 :} an undecayed particle or an unfragmented parton, the latter
being either a single parton or the last one of a parton system.
\iteme{= 2 :} an unfragmented parton, which is followed by more partons
in the same colour-singlet parton system.
\iteme{= 3 :} an unfragmented parton with special colour flow information
stored in \ttt{K(I,4)} and \ttt{K(I,5)}, such that adjacent partons
along the string need not follow each other in the event record.
\iteme{= 4 :} a particle which could have decayed, but did not
within the allowed volume around the original vertex.
\iteme{= 5 :} a particle which is to be forced to decay in the next
\ttt{PYEXEC} call, in the vertex position given (this code is only
set by user intervention).
\iteme{= 11 :} a decayed particle or a fragmented parton, the latter
being either a single parton or the last one of a parton system,
cf. \ttt{= 1}.
\iteme{= 12 :} a fragmented parton, which is followed by more partons
in the same colour-singlet parton system, cf. \ttt{= 2}. Further, a
$\B$ meson which decayed as a $\br{\B}$ one, or vice versa, because of
$\B$--$\br{\B}$ mixing, is marked with this code rather than
\ttt{= 11}.
\iteme{= 13 :} a parton which has been removed when special colour flow
information has been used to rearrange a parton system, cf. \ttt{= 3}.
\iteme{= 14 :} a parton which has branched into further partons, with
special colour-flow information provided, cf. \ttt{= 3}.
\iteme{= 15 :} a particle which has been forced to decay (by user
intervention), cf. \ttt{= 5}.
\iteme{= 21 :} documentation lines used to give a compressed story of
the event at the beginning of the event record.
\iteme{= 31 :} lines with information on sphericity, thrust or cluster
search.
\iteme{= 32 :} tabular output, as generated by \ttt{PYTABU}.
\iteme{= 41 :} a junction, with partons arranged in colour, except that
two quark lines may precede or follow a junction. For instance, a
configuration like $\q_1 \g_1 \, \q_2 \g_2 \,$(junction)$\, \g_3 \q_3$
corresponds to having three strings $\q_1 \g_1$, $\q_2 \g_2$ and
$\q_3 \g_3$ meeting in the junction. The occurence of non-matching colours
easily reveal the $\q_2$ as not being a continuation of the $\q_1 \g_1$
string. Here each $\g$ above is shorthand for an arbitrary number of gluons,
including none. The most general topology allows two junctions in a system,
i.e.\ $\q_1 \g_1 \, \q_2 \g_2 \,$(junction)$\, \g_0\, $(junction)$\, \g_3
\qbar_3 \, \g_4 \qbar_4$. The final $\q/\qbar$ would have status code 1,
the other partons 2. Thus code \ttt{= 41} occurs where \ttt{= 2} would
normally have been used, had the junction been an ordinary parton.
\iteme{= 42 :} a junction, with special colour flow information
stored in \ttt{K(I,4)} and \ttt{K(I,5)}, such that adjacent partons
along the string need not follow each other in the event record.
Thus this code matches the \ttt{= 3} of ordinary partons.
\iteme{= 51 :} a junction of strings which have been fragmented,
cf. \ttt{= 41}. Thus this code matches the \ttt{= 12} of ordinary partons.
\iteme{= 52 :} a junction of strings which have been rearranged in
colour, cf. \ttt{= 42}. Thus this code matches the \ttt{= 13} of ordinary
partons.
\iteme{< 0 :} these codes are never used by the program, and are
therefore usually not affected by operations on the record, such as
\ttt{PYROBO}, \ttt{PYLIST} and event-analysis routines (the exception
is some \ttt{PYEDIT} calls, where lines are moved but not deleted).
Such codes may therefore be useful in some connections.
\end{subentry}
\iteme{K(I,2) :} particle {\KF} code, as described in section
\ref{ss:codes}.
\iteme{K(I,3) :} line number of parent particle, where known,
otherwise 0. Note that the assignment of a particle to a given parton in a
parton system is unphysical, and what is given there is only related to
the way the fragmentation was generated.
\iteme{K(I,4) :} normally the line number of the first daughter;
it is 0 for an undecayed particle or unfragmented parton.
For \ttt{K(I,1) = 3, 13} or \ttt{14}, instead, it contains special
colour-flow information (for internal use only) of the form \\
\ttt{K(I,4)} = 200000000*MCFR + 100000000*MCTO + 10000*ICFR + ICTO, \\
where ICFR and ICTO give the line numbers of the partons from which
the colour comes and to where it goes, respectively; MCFR and
MCTO originally are 0 and are set to 1 when the corresponding
colour connection has been traced in the \ttt{PYPREP} rearrangement
procedure. (The packing may be changed with \ttt{MSTU(5)}.)
The `from' colour position may indicate a parton which branched
to produce the current parton, or a parton created together with
the current parton but with matched anticolour, while the `to'
normally indicates a parton that the current parton branches
into. Thus, for setting up an initial colour configuration, it
is normally only the `from' part that is used, while the `to' part
is added by the program in a subsequent call to parton-shower
evolution (for final-state radiation; it is the other way around
for initial-state radiation). Note that, when using \ttt{PYEVNW} to
generate events, a complementary way of storing the colour flow
information is also used, so-called Les Houches style colour tags
\cite{Boo01}, cf.\ the \ttt{/PYCTAG/} common block.
For \ttt{K(I,1) = 42} or \ttt{52}, see below.
{\bf Note:} normally most users never have to worry about the exact
rules for colour-flow storage, since this is used mainly for
internal purposes. However, when it is necessary to define this
flow, it is recommended to use the \ttt{PYJOIN} routine, since it is
likely that this would reduce the chances of making a mistake.
\iteme{K(I,5) :} normally the line number of the last daughter;
it is 0 for an undecayed particle or unfragmented parton.
For \ttt{K(I,1) = 3, 13} or \ttt{14}, instead, it contains special
colour-flow information (for internal use only) of the form \\
\ttt{K(I,5)} = 200000000*MCFR + 100000000*MCTO + 10000*ICFR + ICTO, \\
where ICFR and ICTO give the line numbers of the partons from which
the anticolour comes and to where it goes, respectively; MCFR
and MCTO originally are 0 and are set to 1 when the corresponding
colour connection has been traced in the \ttt{PYPREP} rearrangement
procedure. For further discussion, see \ttt{K(I,4)}.
For \ttt{K(I,1) = 42} or \ttt{52}, see below.
\iteme{K(I,4), K(I,5) :} For junctions with \ttt{K(I,1) = 42} or \ttt{52}
the colour flow information scheme presented above has to be modified,
since now three colour or anticolour lines meet. Thus the form is\\
\ttt{K(I,4)} = 100000000*MC1 + 10000*\ttt{ITP} + IC1, \\
\ttt{K(I,5)} = 200000000*MC2 + 100000000*MC3 + 10000*IC2 + IC3. \\
The colour flow possibilities are
\begin{entry}
\iteme{ITP = 1 :} junction of three colours in the final state, with
positions as stored in IC1, IC2 and IC3. A typical example would be
neutralino decay to three quarks. Note that the positions need not be
filled by the line numbers of the final quark themselves, but more likely
by the immediate neutralino decay products that thereafter initiate showers
and branch further.
\iteme{ITP = 2 :} junction of three anticolours in the final state,
with positions as stored in IC1, IC2 and IC3.
\iteme{ITP = 3 :} junction of one incoming anticolour to two outgoing
colours, with the anticolour position stored in IC1 and the two colour
ones in IC2 and IC3. A typical example would be an antisquark decaying
to two quarks.
\iteme{ITP = 4 :} junction of one incoming colour to two outgoing
anticolours, with the colour position stored in IC1 and the two anticolour
ones in IC2 and IC3.
\iteme{ITP = 5 :} junction of a colour octet into three colours. The
incoming colour is supposed to pass through unchanged, and so is bookkept
as usual for the particle itself. IC1 is the position of the incoming
anticolour, while IC2 and IC3 are the positions of the new colours
associated with the vanishing of this anticolour. A typical example would
be gluino decay to three quarks.
\iteme{ITP = 6 :} junction of a colour octet into three anticolours. The
incoming anticolour is supposed to pass through unchanged, and so is
bookkept as usual for the particle itself. IC1 is the position of the
incoming colour, while IC2 and IC3 are the positions of the new anticolours
associated with the vanishing of this colour.
\end{entry}
Thus odd (even) \ttt{ITP} code corresponds to a $+1$ ($-1$) change in
baryon number across the junction.\\
The MC1, MC2 and MC3 mark which colour connections have been traced in a
\ttt{PYPREP} rearrangement procedure, as above.
\iteme{P(I,1) :}\label{p:P} $p_x$, momentum in the $x$ direction,
in GeV/$c$.
\iteme{P(I,2) :} $p_y$, momentum in the $y$ direction, in GeV/$c$.
\iteme{P(I,3) :} $p_z$, momentum in the $z$ direction, in GeV/$c$.
\iteme{P(I,4) :} $E$, energy, in GeV.
\iteme{P(I,5) :} $m$, mass, in GeV/$c^2$. In parton showers, with
space-like virtualities, i.e.\ where $Q^2 = - m^2 > 0$,
one puts \ttt{P(I,5)}$ = -Q$.
\iteme{V(I,1) :}\label{p:V} $x$ position of production vertex, in mm.
\iteme{V(I,2) :} $y$ position of production vertex, in mm.
\iteme{V(I,3) :} $z$ position of production vertex, in mm.
\iteme{V(I,4) :} time of production, in mm/$c$
($\approx 3.33 \times 10^{-12}$ s).
\iteme{V(I,5) :} proper lifetime of particle, in mm/$c$
($\approx 3.33 \times 10^{-12}$ s). If the particle is not expected to
decay, \ttt{V(I,5) = 0}. A line with \ttt{K(I,1) = 4}, i.e.\ a
particle that could have decayed, but did not within the
allowed region, has the proper non-zero \ttt{V(I,5)}.
In the absence of electric or magnetic fields, or other
disturbances, the decay vertex \ttt{VP} of an unstable particle
may be calculated as \\
\ttt{VP(j) = V(I,j) + V(I,5)*P(I,j)/P(I,5)},
\ttt{j} = 1--4.
\end{entry}
\subsection{How The Event Record Works}
The event record is the main repository for information about an
event. In the generation chain, it is used as a `scoreboard' for
what has already been done and what remains to do.
This information can be studied by you, to access information
not only about the final state, but also about what came before.
\subsubsection{A simple example}
The first example of section \ref{ss:JETstarted} may help to clarify
what is going on. When \ttt{PY2ENT} is called to generate a $\q\qbar$
pair, the quarks are stored in lines 1 and 2 of the event record,
respectively. Colour information is set to show that they belong
together as a colour singlet. The counter \ttt{N} is also updated
to the value of 2. At no stage is a previously generated event
removed. Lines 1 and 2 are overwritten, but lines 3
onwards still contain whatever may have been there before. This does
not matter, since \ttt{N} indicates where the `real' record ends.
As \ttt{PYEXEC} is called, explicitly by you or indirectly
by \ttt{PY2ENT}, the first entry is considered and found to be
the first parton of a system. Therefore the second entry is also found,
and these two together form a colour singlet parton system, which may
be allowed to fragment. The `string' that fragments is put in line 3
and the fragmentation products in lines 4 through 10 (in this
particular case). At the same time, the $\q$ and $\qbar$ in the
first two lines are marked as having fragmented, and the same for
the string. At this stage, \ttt{N} is 10. Internally in \ttt{PYEXEC}
there is another counter with the value 2, which indicates how far
down in the record the event has been studied.
This second counter is gradually increased by one. If the entry in
the corresponding line can fragment or decay, then fragmentation or
decay is performed.
The fragmentation/decay products are added at the end of the event
record, and \ttt{N} is updated accordingly. The entry is then also
marked as having been treated. For instance, when line 3 is
considered, the `string' entry of this line is seen to have been
fragmented,
and no action is taken. Line 4, a $\rho^+$, is allowed to decay to
$\pi^+ \pi^0$; the decay products are stored in lines 11 and 12,
and line 4 is marked as having decayed. Next, entry 5 is allowed to
decay. The entry in line 6, $\pi^+$, is a stable particle (by
default) and is therefore passed by without any action being taken.
In the beginning of the process, entries are usually unstable, and
\ttt{N} grows faster than the second counter of treated entries.
Later on, an increasing fraction of the entries are stable end
products, and the r\^oles are now reversed, with the second counter
growing faster. When the two coincide, the end of the record has
been reached, and the process can be stopped. All unstable objects
have now been allowed to fragment or decay. They are still present
in the record, so as to simplify the tracing of the history.
Notice that \ttt{PYEXEC} could well be called a second time.
The second counter would then start all over from the beginning, but
slide through until the end without causing any action, since
all objects that can be treated already have been.
Unless some of the relevant switches were changed meanwhile, that
is. For instance, if $\pi^0$ decays were switched off the first time
around but on the second, all the $\pi^0$'s found in the record
would be allowed to decay in the second call. A particle once
decayed is not `undecayed', however, so if the $\pi^0$ is put back
stable and \ttt{PYEXEC} is called a third time, nothing will happen.
\subsubsection{Complete PYTHIA events}
\label{sss:PYrecord}
In a full-blown event generated with {\Py}, the usage of \ttt{PYJETS}
is more complicated, although the general principles survive.
\ttt{PYJETS} is used extensively by many of the generation routines;
indeed it provides the bridge between many of them. The {\Py} event
listing begins (optionally)
with a few lines of event summary, specific to the hard process
simulated and thus not described in the overview above. These
specific parts are covered in the following.
In most instances, only the particles actually produced
are of interest. For \ttt{MSTP(125) = 0}, the event record starts
off with the parton configuration existing after hard interaction,
initial- and final-state radiation, multiple interactions and beam
remnants have been considered. The partons are arranged in colour
singlet clusters, ordered as required for string fragmentation.
Also photons and leptons produced as part of the hard interaction
(e.g.\ from $\q\qbar \to \g \gamma$ or $\u\ubar \to \Z^0 \to \ee$)
appear in this part of the event record. These original entries
appear with pointer \ttt{K(I,3) = 0}, whereas the products of the
subsequent fragmentation and decay have \ttt{K(I,3)} numbers
pointing back to the line of the parent.
The standard documentation, obtained with \ttt{MSTP(125) = 1},
includes a few lines at the beginning of the event record, which
contain a brief summary of the process that has taken place. The
number of lines used depends on the nature of the hard process
and is stored in \ttt{MSTI(4)} for the current event. These lines
all have \ttt{K(I,1) = 21}. For all processes, lines 1 and 2 give
the two incoming particles. When listed with \ttt{PYLIST}, these two
lines will be separated from subsequent ones by a sequence of
`\ttt{======}' signs, to improve readability. For diffractive and
elastic events, the two outgoing states in lines 3 and 4 complete the
list. Otherwise, lines 3 and 4 contain the two partons that initiate
the two initial-state parton showers, and 5 and 6 the end products of
these showers, i.e.\ the partons that enter the hard interaction. With
initial-state radiation switched off, lines 3 and 5 and lines 4 and 6
are identical. For a simple $2 \to 2$ hard scattering, lines 7 and 8 give
the two outgoing partons/particles from the hard interaction, before
any final-state radiation. For $2 \to 2$ processes proceeding via an
intermediate resonance such as $\gammaZ$, $\W^{\pm}$ or $\hrm^0$, the
resonance is found in line 7 and the two outgoing partons/particles in
8 and 9. In some cases one of these may be a resonance in its own
right, or both of them, so that further pairs of lines are added for
subsequent decays. If the decay of a given resonance has been
switched off, then no decay products are listed either in this
initial summary or in the subsequent ordinary listing. Whenever partons
are listed, they are assumed to be on the mass shell for simplicity.
The fact that effective masses may be generated by initial-
and final-state radiation is taken into account in the actual parton
configuration that is allowed to fragment, however. The listing of the
event documentation closes with another line made up of `\ttt{======}'
signs.
A few examples may help clarify the picture. For a single diffractive
event $\p \pbar \to \p_{\mrm{diffr}} \pbar$, the event record will start
with \\
\verb& I K(I,1) K(I,2) K(I,3) & comment \\
\verb& 1 21 2212 0 & incoming $\p$ \\
\verb& 2 21 -2212 0 & incoming $\pbar$ \\
\verb&========================= & not part of record; appears in
listings \\
\verb& 3 21 9902210 1 & outgoing $\p_{\mrm{diffr}}$ \\
\verb& 4 21 -2212 2 & outgoing $\pbar$ \\
\verb&========================= & again not part of record
The typical QCD $2 \to 2$ process would be \\
\verb& I K(I,1) K(I,2) K(I,3) & comment \\
\verb& 1 21 2212 0 & incoming $\p$ \\
\verb& 2 21 -2212 0 & incoming $\pbar$ \\
\verb&========================= & \\
\verb& 3 21 2 1 & $\u$ picked from incoming $\p$ \\
\verb& 4 21 -1 2 & $\dbar$ picked from incoming
$\pbar$ \\
\verb& 5 21 21 3 & $\u$ evolved to $\g$ at hard
scattering \\
\verb& 6 21 -1 4 & still $\dbar$ at hard scattering \\
\verb& 7 21 21 0 & outgoing $\g$ from hard
scattering \\
\verb& 8 21 -1 0 & outgoing $\dbar$ from hard
scattering \\
\verb&========================= &
Note that, where well defined, the \ttt{K(I,3)} code does contain
information as to which side the different partons come from, e.g.
above the gluon in line 5 points back to the $\u$ in line 3,
which points back to the proton in line 1. In the example above, it
would have been possible to associate the scattered g in line 7
with the incoming one in line 5, but this is not possible in the
general case, consider e.g.\ $\g \g \to \g \g$.
A special case is
provided by $\W^+ \W^-$ or $\Z^0 \Z^0$ fusion to an $\hrm^0$. Then the
virtual $\W$'s or $\Z$'s are shown in lines 7 and 8, the $\hrm^0$ in
line 9, and the two recoiling quarks (that emitted the bosons) in 10
and 11, followed by the Higgs decay products. Since the $\W$'s and
$\Z$'s are space-like, what is actually listed as the mass for them
is $-\sqrt{-m^2}$. Thus $\W^+\W^-$ fusion to an $\hrm^0$ in process 8
(not process 124, which is lengthier) might look like \\
\verb& I K(I,1) K(I,2) K(I,3) & comment \\
\verb& 1 21 2212 0 & first incoming $\p$ \\
\verb& 2 21 2212 0 & second incoming $\p$ \\
\verb&========================= & \\
\verb& 3 21 2 1 & $\u$ picked from first $\p$ \\
\verb& 4 21 21 2 & $\g$ picked from second $\p$ \\
\verb& 5 21 2 3 & still $\u$ after initial-state
radiation \\
\verb& 6 21 -4 4 & $\g$ evolved to $\cbar$ \\
\verb& 7 21 24 5 & space-like $\W^+$ emitted by $\u$
quark \\
\verb& 8 21 -24 6 & space-like $\W^-$ emitted by
$\cbar$ quark \\
\verb& 9 21 25 0 & Higgs produced by $\W^+ \W^-$
fusion \\
\verb&10 21 1 5 & $\u$ turned into $\d$ by emission
of $\W^+$ \\
\verb&11 21 -3 6 & $\cbar$ turned into $\sbar$ by
emission of $\W^-$ \\
\verb&12 21 23 9 & first $\Z^0$ coming from decay
of $\hrm^0$ \\
\verb&13 21 23 9 & second $\Z^0$ coming from decay
of $\hrm^0$ \\
\verb&14 21 12 12 & $\nu_{\e}$ from first $\Z^0$
decay \\
\verb&15 21 -12 12 & $\br{\nu}_{\e}$ from first
$\Z^0$ decay \\
\verb&16 21 5 13 & $\b$ quark from second $\Z^0$
decay \\
\verb&17 21 -5 13 & $\bbar$ antiquark from second
$\Z^0$ decay \\
\verb&========================= &
Another special case is when a spectrum of virtual photons are generated
inside a lepton beam, i.e.\ when \ttt{PYINIT} is called with one or
two {\galep} arguments. (Where \textit{lepton} could be either of
\ttt{e-}, \ttt{e+}, \ttt{mu-}, \ttt{mu+}, \ttt{tau-} or \ttt{tau+}.)
Then the documentation section is expanded to reflect the new layer of
administration. Positions 1 and 2 contain the original beam particles,
e.g.\ $\e$ and $\p$ (or $\e^+$ and $\e^-$). In position 3 (and 4 for
$\e^+\e^-$) is (are) the scattered outgoing lepton(s). Thereafter comes
the normal documentation, but starting from the photon rather
than a lepton. For $\e\p$, this means 4 and 5 are the $\gamma^*$
and $\p$, 6 and 7 the shower initiators, 8 and 9 the incoming partons
to the hard interaction, and 10 and 11 the outgoing ones. Thus the
documentation is 3 lines longer (4 for $\e^+\e^-$) than normally.
The documentation lines are often helpful to understand in broad
outline what happened in a given event. However, they only provide
the main points of the process, with many intermediate layers of
parton showers omitted. The documentation can therefore appear internally
inconsistent, if the user does not remember what could have happened
in between. For instance, the listing above would show the Higgs with the
momentum it has before radiation off the two recoiling $\u$ and $\cbar$
quarks is considered. When these showers are included, the Higgs momentum
may shift by the changed recoil. However, this update is not visible in the
initial summary, which thus still shows the Higgs before the showering.
When the Higgs decays, on the other hand, it is the real Higgs momentum
further down in the event record that is used, and that thus sets the
momenta of the decay products that are also copied up to the summary.
Such effects will persist in further decays; e.g. the $\b$ and $\bbar$
shown at the end of the example above are before showers, and may deviate
from the final parton momenta quite significantly. Similar shifts will
also occur e.g. in a $\t \to \b \W^+ \to \b \q \qbar'$ decays,
when the gluon radiation off the $\b$ gives a recoil to the $\W$ that is
not visible in the $\W$ itself but well in its decay products. In summary,
the documentation section should never be mistaken for the physically
observable state in the main section of the event record, and never be
used as part of any realistic event analysis.
(An alternative approach would be in the spirit of the Les Houches Accord
`parton-level' event record, section \ref{ss:PYnewproc}, where the whole
chain of decays normally is carried out before starting the parton showers.
With this approach, one could have an internally consistent summary, but
then in diverging disagreement with the "real" particles after each layer
of shower evolution.)
After these lines with the initial information, the event record looks
the same as for \ttt{MSTP(125) = 0}, i.e.\ first comes the parton
configuration to be fragmented and, after another separator line
`\ttt{======}' in the output (but not the event record), the products
of subsequent fragmentation and decay chains. This ordinary listing
begins in position \ttt{MSTI(4) + 1}. The \ttt{K(I,3)}
pointers for the partons, as well as leptons and photons produced
in the hard interaction, are now pointing towards the documentation
lines above, however. In particular, beam remnants point to 1 or 2,
depending on which side they belong to, and partons emitted in the
initial-state parton showers point to 3 or 4. In the second example
above, the partons produced by final-state radiation will be pointing
back to 7 and 8; as usual, it should be remembered that a specific
assignment to 7 or 8 need not be unique. For the third example,
final-state radiation partons will come both from partons 10 and 11
and from partons 16 and 17, and additionally there will be a
neutrino--antineutrino pair pointing to 14 and 15.
A hadronic event may contain several (semi)hard interactions, when
multiple interactions are allowed. The hardest interaction of an event
is shown in the initial section of the event record, while further
ones are not. Therefore these extra partons, documented in the main
section of the event, do not have a documentation copy to point back to,
and so are assigned \ttt{K(I,3) = 0}.
There exists a third documentation option, \ttt{MSTP(125) = 2}. Here
the history of initial- and final-state parton branchings may be traced,
including all details on colour flow. This information has not been
optimized for user-friendliness, and cannot be recommended for
general usage. With this option, the initial documentation lines
are the same. They are followed by blank lines, \ttt{K(I,1) = 0}, up to
line 100 (can be changed in \ttt{MSTP(126)}). From line 101 onwards
each parton with \ttt{K(I,1) = } 3, 13 or 14 appears with special
colour-flow information in the \ttt{K(I,4)} and \ttt{K(I,5)}
positions. For an ordinary $2 \to 2$ scattering, the two incoming
partons at the hard scattering are stored in lines 101 and 102, and the
two outgoing in 103 and 104. The colour flow between these partons has
to be chosen according to the proper relative probabilities in
cases when many alternatives are possible, see section
\ref{sss:QCDjetclass}.
If there is initial-state radiation, the two partons in lines 101 and
102 are copied down to lines 105 and 106, from which the initial-state
showers are reconstructed backwards step by step. The branching
history may be read by noting that, for a branching $a \to b c$,
the \ttt{K(I,3)} codes of $b$ and $c$ point towards the line number
of $a$. Since the showers are reconstructed backwards, this actually
means that parton $b$ would appear in the listing before parton
$a$ and $c$, and hence have a pointer to a position below
itself in the list. Associated time-like partons $c$ may initiate
time-like showers, as may the partons of the hard scattering. Again
a showering parton or pair of partons will be copied down towards
the end of the list and allowed to undergo successive branchings
$c \to d e$, with $d$ and $e$ pointing towards $c$. The mass of
time-like partons is properly stored in \ttt{P(I,5)}; for space-like
partons $-\sqrt{-m^2}$ is stored instead. After this
section, containing all the branchings, comes the final parton
configuration, properly arranged in colour, followed by all
subsequent fragmentation and decay products, as usual.
\subsection{The HEPEVT Standard}
\label{ss:HEPEVT}
A set of common blocks was developed and agreed on within the
framework of the 1989 LEP physics study, see \cite{Sjo89}.
This standard defines an event record structure which should make
the interfacing of different event generators much simpler.
It would be a major work to rewrite {\Py} to agree with this
standard event record structure. More importantly, the standard
only covers quantities which can be defined unambiguously, i.e.\
which are independent of the particular program used. There are
thus no provisions for the need for colour-flow information in
models based on string fragmentation, etc., so the standard
common blocks would anyway have to be supplemented with additional
event information. The adopted approach is therefore
to retain the \ttt{PYJETS} event record, but supply a routine
\ttt{PYHEPC} which can convert to or from the standard event record.
Owing to a somewhat different content in the two records, some
ambiguities do exist in the translation procedure. \ttt{PYHEPC}
has therefore to be used with some judgement.
In this section, the standard event structure is first presented,
i.e.\ the most important points in \cite{Sjo89} are recapitulated.
Thereafter the conversion routine is described, with particular
attention to ambiguities and limitations.
The standard event record is stored in two common blocks. The second
of these is specifically intended for spin information. Since {\Py}
never (explicitly) makes use of spin information, this latter
common block is not addressed here. A third common block for colour
flow information has been discussed, but never formalized. Note that
a \ttt{CALL PYLIST(5)} can be used to obtain a simple listing of
the more interesting information in the event record.
In order to make the components of the standard more distinguishable
in your programs, the three characters \ttt{HEP} (for High Energy
Physics) have been chosen to be a part of all names.
Originally it was not specified whether real variables should be in
single or double precision. At the time, this meant that single
precision became the default choice, but since then the trend has been
towards increasing precision. In connection with the 1995 LEP~2
workshop, it was therefore agreed to adopt \ttt{DOUBLE PRECISION}
real variables as part of the standard, and also to extend the size
from 2000 to 4000 entries \cite{Kno96}. If, for some reason, one would
want to revert to single precision, this would only require trivial
changes to the code of the \ttt{PYHEPC} conversion routine described
below.
\drawboxfour{~PARAMETER (NMXHEP=4000)}
{~COMMON/HEPEVT/NEVHEP,NHEP,ISTHEP(NMXHEP),IDHEP(NMXHEP),}
{\&JMOHEP(2,NMXHEP),JDAHEP(2,NMXHEP),PHEP(5,NMXHEP),VHEP(4,NMXHEP)}
{~DOUBLE PRECISION PHEP, VHEP}
\label{p:HEPEVT}\begin{entry}
\itemc{Purpose:} to contain an event record in a
Monte Carlo-independent format.
\iteme{NMXHEP:} maximum numbers of entries (particles) that can
be stored in the common block. The default value of 4000 can be changed
via the parameter construction. In the translation, it is
checked that this value is not exceeded.
\iteme{NEVHEP:} is normally the event number, but may have special
meanings, according to the description below:
\begin{subentry}
\iteme{> 0 :} event number, sequentially increased by 1 for each call
to the main event generation routine, starting with 1 for the
first event generated.
\iteme{= 0 :} for a program which does not keep track of event numbers,
as some of the {\Py} routines.
\iteme{= -1 :} special initialization record; not used by {\Py}.
\iteme{= -2 :} special final record; not used by {\Py}.
\end{subentry}
\iteme{NHEP:} the actual number of entries stored in the current event.
These are found in the first \ttt{NHEP} positions of the respective
arrays below. Index \ttt{IHEP}, 1 $\leq$ \ttt{IHEP} $\leq$ \ttt{NHEP},
is used below to denote a given entry.
\iteme{ISTHEP(IHEP):} status code for entry \ttt{IHEP}, with the
following meanings:
\begin{subentry}
\iteme{= 0 :} null entry.
\iteme{= 1 :} an existing entry, which has not decayed or fragmented.
This is the main class of entries, which represents the
`final state' given by the generator.
\iteme{= 2 :} an entry which has decayed or fragmented and is
therefore not appearing in the final state, but is retained for
event history information.
\iteme{= 3 :} a documentation line, defined separately from the event
history. This could include the two incoming reacting particles, etc.
\iteme{= 4 - 10 :} undefined, but reserved for future standards.
\iteme{= 11 - 200 :} at the disposal of each model builder for
constructs specific to his program, but equivalent to a null line
in the context of any other program.
\iteme{= 201 - :} at the disposal of users, in particular for event
tracking in the detector.
\end{subentry}
\iteme{IDHEP(IHEP) :} particle identity, according to the PDG
standard. The four additional codes 91--94 have been introduced
to make the event history more legible, see section \ref{ss:codes}
and the \ttt{MSTU(16)} description of how daughters can point back
to them.
\iteme{JMOHEP(1,IHEP) :} pointer to the position where the mother
is stored. The value is 0 for initial entries.
\iteme{JMOHEP(2,IHEP) :} pointer to position of second mother.
Normally only one mother exists, in which case the value 0 is to be
used. In {\Py}, entries with codes 91--94 are the only ones to have
two mothers. The flavour contents of these objects, as well as
details of momentum sharing, have to be found by looking at the
mother partons, i.e.\ the two partons in positions \ttt{JMOHEP(1,IHEP)}
and \ttt{JMOHEP(2,IHEP)} for a cluster or a shower system, and the
range
\ttt{JMOHEP(1,IHEP)}--\ttt{JMOHEP(2,IHEP)} for a string or an
independent fragmentation parton system.
\iteme{JDAHEP(1,IHEP) :} pointer to the position of the first daughter.
If an entry has not decayed, this is 0.
\iteme{JDAHEP(2,IHEP) :} pointer to the position of the last daughter.
If an entry has not decayed, this is 0. It is assumed that daughters are
stored sequentially, so that the whole range
\ttt{JDAHEP(1,IHEP)}--\ttt{JDAHEP(2,IHEP)} contains daughters. This
variable should be set also when only one daughter is present, as in
$\K^0 \to \K_{\mrm{S}}^0$ decays, so that looping from the first
daughter to the last one works transparently.
Normally daughters are stored after mothers, but in backwards
evolution of initial-state radiation the opposite may appear,
i.e.\ that mothers are found below the daughters they branch into.
Also, the two daughters then need not appear one after the other,
but may be separated in the event record.
\iteme{PHEP(1,IHEP) :} momentum in the $x$ direction, in GeV/$c$.
\iteme{PHEP(2,IHEP) :} momentum in the $y$ direction, in GeV/$c$.
\iteme{PHEP(3,IHEP) :} momentum in the $z$ direction, in GeV/$c$.
\iteme{PHEP(4,IHEP) :} energy, in GeV.
\iteme{PHEP(5,IHEP) :} mass, in GeV/$c^2$. For space-like partons,
it is allowed to use a negative mass, according to
\ttt{PHEP(5,IHEP)}$ = -\sqrt{-m^2}$.
\iteme{VHEP(1,IHEP) :} production vertex $x$ position, in mm.
\iteme{VHEP(2,IHEP) :} production vertex $y$ position, in mm.
\iteme{VHEP(3,IHEP) :} production vertex $z$ position, in mm.
\iteme{VHEP(4,IHEP) :} production time, in mm/$c$
($\approx 3.33 \times 10^{-12}$ s).
\end{entry}
\boxsep
This completes the brief description of the standard. In {\Py}, the
routine \ttt{PYHEPC} is provided as an interface.
\drawbox{CALL PYHEPC(MCONV)}\label{p:PYHEPC}
\begin{entry}
\itemc{Purpose:} to convert between the \ttt{PYJETS} event record and
the \ttt{HEPEVT} event record.
\iteme{MCONV :} direction of conversion.
\begin{subentry}
\iteme{= 1 :} translates the current \ttt{PYJETS} record into the
\ttt{HEPEVT} one, while leaving the original \ttt{PYJETS} one
unaffected.
\iteme{= 2 :} translates the current \ttt{HEPEVT} record into the
\ttt{PYJETS} one, while leaving the original \ttt{HEPEVT} one
unaffected.
\end{subentry}
\end{entry}
\boxsep
The conversion of momenta is trivial: it is just a matter of exchanging
the order of the indices. The vertex information is but little more
complicated; the extra fifth component present in \ttt{PYJETS} can be
easily
reconstructed from other information for particles which have decayed.
(Some of the advanced features made possible by this component, such as
the possibility to consider decays within expanding spatial volumes in
subsequent \ttt{PYEXEC} calls, cannot be used if the record is
translated back and forth, however.) Also, the particle codes
\ttt{K(I,2)} and \ttt{IDHEP(I)}
are identical, since they are both based on the PDG codes.
The remaining, non-trivial areas deal with the status codes and the
event history. In moving from \ttt{PYJETS} to \ttt{HEPEVT},
information on colour flow is lost. On the other hand, the position
of a second mother, if any, has to be found; this only affects lines
with \ttt{K(I,2) =} 91--94. Also, for lines with \ttt{K(I,1) = } 13
or 14, the daughter pointers have to be found. By and large,
however, the translation from \ttt{PYJETS} to \ttt{HEPEVT}
should cause little problem, and there should never be any need for
user intervention. (We assume that {\Py} is run with the default
\ttt{MSTU(16) = 1} mother pointer assignments, otherwise some
discrepancies with respect to the proposed standard event history
description will be present.)
In moving from \ttt{HEPEVT} to \ttt{PYJETS}, information on a second
mother is lost. Any codes \ttt{ISTHEP(I)} not equal to 1, 2 or 3 are
translated into \ttt{K(I,1) = 0}, and so all entries with
\ttt{K(I,1)} $\geq 30$ are effectively lost in a translation back and
forth. All entries with \ttt{ISTHEP(I) = 2} are translated
into \ttt{K(I,1) = 11}, and so entries of type
\ttt{K(I,1) = 12, 13, 14} or \ttt{15} are never found. There is thus
no colour-flow information available for partons which have
fragmented. For partons with \ttt{ISTHEP(I) = 1},
i.e.\ which have not fragmented, an attempt is made to subdivide the
partonic system into colour singlets, as required for subsequent
string fragmentation. To this end, it is assumed that partons are
stored sequentially along strings. Normally, a string would then start
at a $\q$ ($\qbar$) or $\qbar\qbar$ ($\q\q$) entry, cover a number
of intermediate gluons, and end at a $\qbar$ ($\q$) or $\q\q$
($\qbar\qbar$) entry. Particles could be interspersed in this list with
no adverse effects, i.e.\ a $\u-\g-\gamma-\ubar$
sequence would be interpreted as a $\u-\g-\ubar$ string plus an
additional photon. A closed gluon loop would be assumed to be made up
of a sequential listing of the gluons, with the string continuing from
the last gluon up back to the first one. Contrary to the previous, open
string case, the appearance of any particle but a gluon would therefore
signal the end of the gluon loop. For example, a $\g-\g-\g-\g$ sequence
would be interpreted as one single four-gluon loop, while a
$\g-\g-\gamma-\g-\g$ sequence would be seen as composed of two
2-gluon systems.
If these interpretations, which are not unique, are not to your liking,
it is up to you to correct them, e.g.\ by using
\ttt{PYJOIN} to tell exactly which partons should be joined,
in which sequence, to give a string. Calls to \ttt{PYJOIN}
(or the equivalent) are also necessary if \ttt{PYSHOW} is to be used
to have some partons develop a shower.
For practical applications, one should note that $\ee$ events,
which have been allowed to shower but not to fragment, do have partons
arranged in the order assumed above, so that a translation to
\ttt{HEPEVT} and back does not destroy the possibility to perform
fragmentation by a simple \ttt{PYEXEC} call. Also the hard interactions
in hadronic events fulfil this condition, while problems may appear in the
multiple interaction scenario, where several closed $\g\g$ loops may
appear directly following one another, and thus would be
interpreted as a single multigluon loop after translation back and
forth.
\clearpage
\section{The Old Electron--Positron Annihilation Routines}
\label{s:JETSETproc}
{}From the {\Je} package, {\Py} inherits routines for the dedicated
simulation of two hard processes in $\e^+\e^-$ annihilation.
The process of main interest is $\ee \to \gammaZ \to \q \qbar$.
The description provided by the \ttt{PYEEVT} routine has been a main
staple from PETRA days up to the LEP1 era. Nowadays it is superseded
by process 1 of the main {\Py} event generation machinery, see section
\ref{sss:WZclass}. This latter process offers a better description of
flavour selection, resonance shape and initial-state radiation. It can
also, optionally, be used with the second-order matrix element machinery
documented in this section. For backwards compatibility, however, the
old routines have still been retained here. There are also a few
features found in the routines in this section, and not in the other
ones, such as polarized incoming beams.
For the process $\ee \to \gammaZ \to \q \qbar$,
higher-order QCD corrections can be obtained either with parton
showers or with second-order matrix elements. The details of the
parton-shower evolution are given in section \ref{s:showinfi},
while this section contains the matrix-element description, including
a summary of the older algorithm for initial-state photon radiation
used here.
The other standalone hard process in this section is $\Upsilon$ decay to
$\g \g \g$ or $\gamma \g \g$, which is briefly commented on.
The main sources of information for this chapter are
refs. \cite{Sjo83,Sjo86,Sjo89}.
\subsection{Annihilation Events in the Continuum}
\label{ss:eematrix}
The description of $\ee$ annihilation into hadronic events involves a
number of components: the $s$ dependence of the total cross section
and flavour composition, multiparton matrix elements, angular
orientation of events, initial-state photon bremsstrahlung
and effects of initial-state electron polarization.
Many of the published formulae
have been derived for the case of massless outgoing quarks. For each
of the components described in the following, we will begin by
discussing the massless case, and then comment on what is done to
accommodate massive quarks.
\subsubsection{Electroweak cross sections}
In the Standard Model, fermions have the following couplings
(illustrated here for the first generation):
\begin{center}
\begin{tabular}{lll}
$e_{\nu} = 0$, & $v_{\nu} = 1$, & $a_{\nu} = 1$, \\
$e_{\e} = -1$, & $v_{\e} = -1 + 4\ssintw$, & $a_{\e} = -1$, \\
$e_{\u} = 2/3$, & $v_{\u} = 1 - 8\ssintw /3$, & $a_{\nu} = 1$, \\
$e_{\d} = -1/3$, & $v_{\d} = -1 + 4\ssintw /3$, & $a_{\d} = -1$, \\
\end{tabular}
\end{center}
with $e$ the electric charge, and $v$ and $a$ the vector and axial
couplings to the $\Z^0$. The relative energy dependence of the weak
neutral current to the electromagnetic one is given by
\begin{equation}
\chi(s) = \frac{1}{16\ssintw\scostw} \;
\frac{s}{s - m_{\Z}^2 + i m_{\Z}\Gamma_{\Z}} ~,
\label{ee:chis}
\end{equation}
where $s = E_{\mrm{cm}}^2$.
In this section the electroweak mixing parameter $\ssintw$ and the
$\Z^0$ mass $m_{\Z}$ and width $\Gamma_{\Z}$ are considered as
constants to be given by you (while the full {\Py} event generation
machinery itself calculates an $s$-dependent width).
Although the incoming $\e^+$ and $\e^-$ beams are normally
unpolarized, we have included the possibility of polarized beams,
following the formalism of \cite{Ols80}. Thus the incoming
$\e^+$ and $\e^-$ are characterized by polarizations
$\mbf{P}^{\pm}$ in the rest frame of the particles:
\begin{equation}
\mbf{P}^{\pm} = P_{\mrm{T}}^{\pm} \hat{\mbf{s}}^{\pm} +
P_{\mrm{L}}^{\pm} \hat{\mbf{p}}^{\pm} ~,
\end{equation}
where $0 \leq P_{\mrm{T}}^{\pm} \leq 1$ and
$-1 \leq P_{\mrm{L}}^{\pm} \leq 1$, with the constraint
\begin{equation}
(\mbf{P}^{\pm})^2 = (P_{\mrm{T}}^{\pm})^2 + (P_{\mrm{L}}^{\pm})^2
\leq 1 ~.
\end{equation}
Here $\hat{\mbf{s}}^{\pm}$ are unit vectors perpendicular to the beam
directions $\hat{\mbf{p}}^{\pm}$. To be specific, we choose a
right-handed coordinate frame with
$\hat{\mbf{p}}^{\pm} = (0,0, \mp 1)$,
and standard transverse polarization directions (out of the machine
plane for storage rings) $\hat{\mbf{s}}^{\pm} = (0, \pm 1,0)$, the
latter corresponding to azimuthal angles $\varphi^{\pm} = \pm \pi /2$.
As free parameters in the program we choose $P_{\mrm{L}}^+$,
$P_{\mrm{L}}^-$, $P_{\mrm{T}} = \sqrt{P_{\mrm{T}}^+ P_{\mrm{T}}^-}$
and $\Delta \varphi = (\varphi^+ + \varphi^-) /2$.
In the massless QED case, the probability to produce a flavour $\f$ is
proportional to $e_{\f}^2$, i.e up-type quarks are four times as likely
as down-type ones. In lowest-order massless QFD (Quantum Flavour Dynamics;
part of the Standard Model) the corresponding
relative probabilities are given by \cite{Ols80}
\begin{eqnarray}
h_{\f}(s) & = & e_{\e}^2 \, (1 - P_{\mrm{L}}^+ P_{\mrm{L}}^-)
\, e_{\f}^2 \, + \, 2 e_{\e} \left\{ v_{\e}
(1 - P_{\mrm{L}}^+ P_{\mrm{L}}^-) - a_{\e}
(P_{\mrm{L}}^- - P_{\mrm{L}}^+)
\right\} \, \Re\mrm{e}\chi(s) \, e_{\f} v_{\f} \, + \nonumber \\
& & + \, \left\{ (v_{\e}^2 + a_{\e}^2) (1 - P_{\mrm{L}}^+
P_{\mrm{L}}^-) - 2 v_{\e} a_{\e} (P_{\mrm{L}}^- - P_{\mrm{L}}^+)
\right\} \,
\left| \chi(s) \right|^2 \, \left\{ v_{\f}^2 + a_{\f}^2 \right\} ~,
\label{ee:hf}
\end{eqnarray}
where $\Re\mrm{e}\chi(s)$ denotes the real part of $\chi(s)$.
The $h_{\f}(s)$ expression depends both on the $s$ value and on the
longitudinal polarization of the $\e^{\pm}$ beams in a non-trivial way.
The cross section for the process $\ee \to \gammaZ \to \f \fbar$
may now be written as
\begin{equation}
\sigma_{\f}(s) = \frac{4 \pi \alphaem^2}{3 s} R_{\f}(s) ~,
\end{equation}
where $R_{\f}$ gives the ratio to the lowest-order QED cross section for
the process $\ee \to \mu^+ \mu^-$,
\begin{equation}
R_{\f}(s) = N_C \, R_{\mrm{QCD}} \, h_{\f}(s) ~.
\end{equation}
The factor of $N_C = 3$ counts the number of colour states available
for the $\q\qbar$ pair. The $R_{\mrm{QCD}}$
factor takes into account QCD loop corrections to the cross section.
For $n_f$ effective flavours (normally $n_f =5$)
\begin{equation}
R_{\mrm{QCD}} \approx 1 + \frac{\alphas}{\pi} + (1.986 - 0.115 n_f)
\left( \frac{\alphas}{\pi} \right)^2 + \cdots
\label{ee:RQCD}
\end{equation}
in the $\br{\mrm{MS}}$ renormalization scheme \cite{Din79}.
Note that $R_{\mrm{QCD}}$ does not affect the relative quark-flavour
composition, and so is of peripheral interest here.
(For leptons the $N_C$ and $R_{\mrm{QCD}}$ factors would be absent,
i.e.\ $N_C \, R_{\mrm{QCD}} = 1$, but leptonic final states are not
generated by this routine.)
Neglecting higher-order QCD and QFD effects, the corrections for
massive quarks are given in terms of the velocity $\beta_{\f}$ of a
fermion with mass $m_{\f}$, $\beta_{\f} = \sqrt{ 1 - 4 m_{\f}^2 /s}$,
as follows. The vector quark current terms in $h_{\f}$ (proportional to
$e_{\f}^2$, $e_{\f} v_{\f}$, or $v_{\f}^2$) are multiplied by a
threshold factor $\beta_{\f} (3 - \beta_{\f}^2) /2$, while the axial
vector quark current term (proportional to $a_{\f}^2$) is
multiplied by $\beta_{\f}^3$. While inclusion of quark masses in the
QFD formulae decreases the total cross section, first-order QCD
corrections tend in the opposite direction \cite{Jer81}. Na\"{\i}vely,
one would expect one factor of $\beta_{\f}$ to get cancelled. So far,
the available options are either to include threshold factors
in full or not at all.
Given that all five quarks are light at
the scale of the $\Z^0$, the issue of quark masses is not really
of interest at LEP. Here, however, purely weak corrections are
important, in particular since they change the $\b$ quark
partial width differently from that of the other ones \cite{Kuh89}.
No such effects are included in the program.
\subsubsection{First-order QCD matrix elements}
The Born process $\ee \to \q \qbar$ is modified in first-order
QCD by the probability for the $\q$ or $\qbar$ to
radiate a gluon, i.e.\ by the process $\ee \to \q \qbar \g$.
The matrix element is conveniently given in terms of scaled energy
variables in the c.m.\ frame of the event,
$x_1 = 2E_{\q}/E_{\mrm{cm}}$,
$x_2 = 2E_{\qbar}/E_{\mrm{cm}}$,
and $x_3 = 2E_{\g}/E_{\mrm{cm}}$,
i.e.\ $x_1 + x_2 + x_3 = 2$. For massless
quarks the matrix element reads \cite{Ell76}
\begin{equation}
\frac{1}{\sigma_0} \, \frac{\d \sigma}{\d x_1 \, \d x_2} =
\frac{\alphas}{2\pi} \, C_F \,
\frac{x_1^2 + x_2^2}{(1-x_1)(1-x_2)} ~,
\label{ee:ME3j}
\end{equation}
where $\sigma_0$ is the lowest-order cross section, $C_F = 4/3$ is the
appropriate colour factor, and
the kinematically allowed region is $0 \leq x_i \leq 1, i = 1, 2, 3$.
By kinematics, the $x_k$ variable for parton $k$ is related to the
invariant mass $m_{ij}$ of the other two partons $i$ and $j$ by
$y_{ij} = m_{ij}^2/E_{\mrm{cm}}^2 = 1 - x_k$.
The strong coupling constant $\alphas$ is in first order given by
\begin{equation}
\alphas(Q^2) = \frac{12\pi}{(33-2n_f) \, \ln(Q^2/\Lambda^2)} ~.
\label{ee:aS3j}
\end{equation}
Conventionally $Q^2 = s = E_{\mrm{cm}}^2$; we will return to this
issue below.
The number of flavours $n_f$ is 5 for LEP applications, and so the
$\Lambda$ value determined is $\Lambda_5$ (while e.g.\ most
Deeply Inelastic Scattering studies refer to $\Lambda_4$,
the $Q^2$ scales for these experiments historically having been
below the bottom threshold).
The $\alphas$ values are matched at flavour thresholds, i.e.\
as $n_f$ is changed the $\Lambda$ value is also changed. It is
therefore the derivative of $\alphas$ that changes at a
threshold, not $\alphas$ itself.
In order to separate 2-jets from 3-jets, it is useful to
introduce jet-resolution parameters. This can be done in several
different ways. Most famous are the $y$ and $(\epsilon, \delta)$
procedures. We will only refer to the $y$ cut, which is the one
used in the program. Here a 3-parton configuration is called
a 2-jet event if
\begin{equation}
\min_{i,j} (y_{ij}) = \min_{i,j} \left( \frac{m_{ij}^2}{E_{\mrm{cm}}^2}
\right) < y ~.
\end{equation}
The cross section in eq.~(\ref{ee:ME3j}) diverges for
$x_1 \rightarrow 1$ or $x_2 \rightarrow 1$ but, when
first-order propagator and vertex corrections are included,
a corresponding singularity with opposite sign appears in the
$\q \qbar$ cross section, so that the total cross section is finite.
In analytical calculations, the average value of any well-behaved
quantity ${\cal Q}$ can therefore be calculated as
\begin{equation}
\left\langle {\cal Q} \right\rangle =
\frac{1}{\sigma_{\mrm{tot}}} \lim_{y \rightarrow 0}
\left( {\cal Q}(\mrm{2parton}) \, \sigma_{\mrm{2parton}}(y) +
\int_{y_{ij} > y} {\cal Q}(x_1,x_2) \,
\frac{\d \sigma_{\mrm{3parton}}}{\d x_1 \, \d x_2} \,
\d x_1 \, \d x_2 \right) ~,
\label{ee:Obs}
\end{equation}
where any explicit $y$ dependence disappears in the limit
$y \rightarrow 0$.
In a Monte Carlo program, it is not possible to
work with a negative total 2-jet rate, and thus it is necessary to
introduce a fixed non-vanishing $y$ cut in the 3-jet
phase space. Experimentally, there is evidence for the need of a
low $y$ cut, i.e.\ a large 3-jet rate.
For LEP applications, the recommended value is $y = 0.01$,
which is about as far down as one can go and still retain a positive
2-jet rate. With $\alphas = 0.12$, in full second-order QCD
(see below), the $2:3:4$ jet composition is then approximately
$11 \% : 77 \% : 12 \%$. Since $\alphas$ varies only slowly with
energy, it is not possible to go much below $y = 0.01$ even at
future Linear Collider energies.
Note, however, that initial-state QED radiation may
occasionally lower the c.m.\ energy significantly, i.e.\ increase
$\alphas$, and thereby bring the 3-jet fraction above unity
if $y$ is kept fixed at 0.01 also in those events. Therefore,
at PETRA/PEP energies, $y$ values slightly above 0.01 are needed.
In addition to the $y$ cut, the program contains a cut on the
invariant mass $m_{ij}$ between any two partons, which is typically
required to be larger than 2 GeV. This cut corresponds to the
actual merging of two nearby parton jets, i.e.\ where a treatment with
two separate partons rather than one would be superfluous in view
of the smearing arising from the subsequent fragmentation. Since
the cut-off mass scale $\sqrt{y} E_{\mrm{cm}}$ normally is much larger,
this additional cut only enters for events at low energies.
For massive quarks, the amount of QCD radiation is slightly reduced
\cite{Iof78}:
\begin{eqnarray}
\frac{1}{\sigma_0} \, \frac{\d \sigma}{\d x_1 \, \d x_2} & = &
\frac{\alphas}{2\pi}
\, C_F \, \left\{ \frac{x_1^2 + x_2^2}{(1-x_1)(1-x_2)} -
\frac{4 m_{\q}^2}{s} \left( \frac{1}{1-x_1} + \frac{1}{1-x_2}
\right) \right. \nonumber \\[1mm]
& & - \left. \frac{2 m_{\q}^2}{s} \left( \frac{1}{(1-x_1)^2} +
\frac{1}{(1-x_2)^2} \right) - \frac{4 m_{\q}^4}{s^2}
\left( \frac{1}{1-x_1} + \frac{1}{1-x_2} \right)^2 \right\} ~.
\label{ee:threejMEmass}
\end{eqnarray}
Properly, the above expression is only valid for the vector part of the
cross section, with a slightly different expression for the axial part,
but here the one above is used for it all.
In addition, the phase space for emission is reduced by the
requirement
\begin{equation}
\frac{(1-x_1)(1-x_2)(1-x_3)}{x_3^2} \geq \frac{m_{\q}^2}{s} ~.
\end{equation}
For $\b$ quarks at LEP energies, these corrections are fairly small.
\subsubsection{Four-jet matrix elements}
Two new event types are added in second-order QCD,
$\ee \to \q \qbar \g \g$ and $\ee \to \q \qbar \q' \qbar'$.
The 4-jet cross section has been calculated by several
groups \cite{Ali80a,Gae80,Ell81,Dan82}, which agree on the result.
The formulae are too lengthy to be quoted here. In one of the
calculations \cite{Ali80a}, quark masses were explicitly included,
but here only the massless expressions are included, as taken
from \cite{Ell81}. Here the angular orientation of the event has been
integrated out, so that five independent internal kinematical
variables remain. These may be related to the six $y_{ij}$ and
the four $y_{ijk}$ variables,
$y_{ij} = m_{ij}^2 / s = (p_i + p_j)^2 / s$ and
$y_{ijk} = m_{ijk}^2 / s = (p_i + p_j + p_k)^2 / s$,
in terms of which the matrix elements are given.
The original calculations were for the pure $\gamma$-exchange case;
it has been pointed out \cite{Kni89} that an additional
contribution to the $\ee \to \q \qbar \q' \qbar'$ cross section
arises from the axial part of the $\Z^0$. This term is not included
in the program, but fortunately it is finite and small.
Whereas the way the string, i.e.\ the fragmenting colour flux tube,
is stretched is uniquely given in $\q \qbar \g$ event, for
$\q \qbar \g \g$ events there are two possibilities:
\mbox{$\q - \g_1 - \g_2 - \qbar$} or
\mbox{$\q - \g_2 - \g_1 - \qbar$}.
A knowledge of quark and gluon colours, obtained by perturbation
theory, will uniquely specify the stretching of the string, as long
as the two gluons do not have the same colour. The probability for
the latter is down in magnitude by a factor $1 / N_C^2 = 1 / 9$.
One may either choose to neglect these terms entirely, or to keep
them for the choice of kinematical setup, but then drop them at the
choice of string drawing \cite{Gus82}. We have adopted the latter
procedure. Comparing the two possibilities, differences are
typically 10--20\% for a given kinematical configuration,
and less for the total 4-jet cross section, so from a practical
point of view this is not a major problem.
In higher orders, results depend on the renormalization scheme;
we will use $\br{\mrm{MS}}$ throughout. In addition to this choice,
several possible forms can be chosen for $\alphas$,
all of which are equivalent to that order but differ in higher
orders. We have picked the recommended standard \cite{PDG88}
\begin{equation}
\label{ee:aS4j}
\alphas(Q^2) =
\frac{12\pi}{(33-2n_f) \, \ln (Q^2 / \Lambda^2_{\br{\mrm{MS}}})}
\left\{ 1 - 6 \, \frac{153-19n_f}{(33-2n_f)^2} \,
\frac{\ln (\ln ( Q^2 / \Lambda^2_{\br{\mrm{MS}}}))}
{\ln ( Q^2 / \Lambda^2_{\br{\mrm{MS}}})}
\right\} ~.
\end{equation}
\subsubsection{Second-order three-jet matrix elements}
As for first order, a full second-order calculation consists both of
real parton emission terms and of vertex and propagator corrections.
These modify the 3-jet and 2-jet cross sections.
Although there was some initial confusion, everybody soon agreed
on the size of the loop corrections \cite{Ell81,Ver81,Fab82}.
In analytic calculations, the procedure of eq.~(\ref{ee:Obs}),
suitably expanded, can therefore be used unambiguously for a
well-behaved variable.
For Monte Carlo event simulation, it is again necessary to impose
some finite jet-resolution criterion. This means that four-parton
events which fail the cuts should be reassigned either to the
3-jet or to the 2-jet event class. It is this area that
caused quite a lot of confusion in the past
\cite{Kun81,Got82,Ali82,Zhu83,Gut84,Gut87,Kra88},
and where full agreement does not exist. Most likely, agreement
will never be reached, since there are indeed ambiguous points
in the procedure, related to uncertainties on the theoretical
side, as follows.
For the $y$-cut case, any two partons with an invariant mass
$m_{ij}^2 < y E_{\mrm{cm}}^2$ should be recombined into one. If the
four-momenta are simply added, the sum will correspond to a parton
with a positive mass, namely the original $m_{ij}$.
The loop corrections are given in terms of final
massless partons, however. In order to perform the (partial)
cancellation between the four-parton real and the 3-parton
virtual contributions, it is therefore necessary to get rid of
the bothersome mass in the four-parton states. Several
recombinations are used in practice, which go under names such as
`E', `E0', `p' and `p0' \cite{OPA91}. In the `E'-type schemes,
the energy of a recombined parton is given by $E_{ij} = E_i + E_j$,
and three-momenta may have to be adjusted accordingly. In the
`p'-type schemes, on the other hand, three-momenta are added,
$\mbf{p}_{ij} = \mbf{p}_i + \mbf{p}_j$, and then energies may have
to be adjusted. These procedures result in different 3-jet
topologies, and therefore in different second-order differential
3-jet cross sections.
Within each scheme, a number of lesser points remain to be dealt
with, in particular what to do if a recombination of a nearby parton
pair were to give an event with a non-$\q\qbar\g$ flavour structure.
This code contains two alternative second-order 3-jet
implementations, the GKS and the ERT(Zhu) ones. The latter is the
recommended one
and default. Other parameterizations have also been made
available that run together with {\Je}~6 (but not adopted to the
current program), see \cite{Sjo89,Mag89}.
The GKS option is based on the GKS \cite{Gut84} calculation, where
some of the original mistakes in FKSS \cite{Fab82} have been
corrected. The GKS formulae have the advantage of giving the
second-order corrections in closed analytic form, as not-too-long
functions of $x_1$, $x_2$, and the $y$ cut. However, it is
today recognized, also by the authors, that important
terms are still missing, and that the matrix elements
should therefore not be taken too seriously. The option is thus
kept mainly for backwards compatibility.
The ERT(Zhu) generator \cite{Zhu83} is based on the ERT matrix elements
\cite{Ell81}, with a Monte Carlo recombination procedure suggested
by Kunszt \cite{Kun81} and developed by Ali \cite{Ali82}. It has
the merit of giving corrections in a convenient, parameterized form.
For practical applications, the main limitation is that the
corrections are only given for discrete values of the cut-off
parameter $y$, namely $y$ = 0.01, 0.02, 0.03, 0.04, and 0.05.
At these $y$ values, the full second-order 3-jet cross section is
written in terms of the `ratio function' $R(X,Y;y)$, defined by
\begin{equation}
\frac{1}{\sigma_0} \frac{\d \sigma_3^{\mrm{tot}}}{\d X \, \d Y} =
\frac{\alphas}{\pi} A_0(X,Y)
\left\{ 1 + \frac{\alphas}{\pi} R(X,Y;y) \right\} ~,
\label{ee:Zhupar}
\end{equation}
where $X = x_1 - x_2 = x_{\q} - x_{\qbar}$, $Y = x_3 = x_g$,
$\sigma_0$ is the lowest-order hadronic cross section,
and $A_0(X,Y)$ the standard first-order 3-jet cross section,
cf. eq.~(\ref{ee:ME3j}).
By Monte Carlo integration, the value of $R(X,Y;y)$ is
evaluated in bins of $(X,Y)$, and the result parameterized
by a simple function $F(X,Y;y)$. Further details are found in
\cite{Sjo89}.
\subsubsection{The matrix-element event generator scheme}
The program contains parameterizations, separately, of the total
first-order 3-jet rate, the total second-order 3-jet rate,
and the total 4-jet rate, all as functions of $y$ (with
$\alphas$ as a separate prefactor).
These parameterizations have been obtained as follows:
\begin{Itemize}
\item
The first-order 3-jet matrix element is almost analytically
integrable; some small finite pieces were obtained by a truncated
series expansion of the relevant integrand.
\item
The GKS second-order 3-jet matrix elements were integrated for
40 different $y$-cut values, evenly distributed in $\ln y$ between
a smallest value $y = 0.001$ and the kinematical limit $y = 1/3$.
For each $y$ value, 250\,000 phase-space points were generated,
evenly in $\d \ln (1-x_i) = \d x_i/(1-x_i)$, $i = 1,2$, and the
second-order 3-jet rate in the point evaluated. The properly
normalized sum of weights in each of the 40 $y$ points were
then fitted to a polynomial in $\ln(y^{-1}-2)$. For the ERT(Zhu)
matrix elements the parameterizations in eq.~(\ref{ee:Zhupar})
were used to perform a corresponding Monte Carlo integration for
the five $y$ values available.
\item
The 4-jet rate was integrated numerically, separately for
$\q \qbar \g \g$ and $\q \qbar \q' \qbar'$ events, by generating large
samples of 4-jet phase-space points
within the boundary $y = 0.001$. Each point was classified according
to the actual minimum $y$ between any two partons. The same
events could then be used to update the summed weights for 40
different counters, corresponding to $y$ values evenly distributed
in $\ln y$ between $y = 0.001$ and the kinematical limit $y = 1/6$.
In fact, since
the weight sums for large $y$ values only received contributions
from few phase-space points, extra (smaller) subsamples of events were
generated with larger $y$ cuts. The summed weights,
properly normalized, were then parameterized in terms of
polynomials in $\ln(y^{-1} - 5)$.
Since it turned out to be difficult to obtain one single good fit
over the whole range of $y$ values, different parameterizations are
used above and below $y=0.018$. As originally given, the
$\q \qbar \q' \qbar'$ parameterization only took into account four
$\q'$ flavours, i.e.\ secondary $\b \bbar$ pairs were not generated,
but this has been corrected for LEP.
\end{Itemize}
In the generation stage, each event is treated on its own, which means
that the $\alphas$ and $y$ values may be allowed to vary from event to
event. The main steps are the following.
\begin{Enumerate}
\item
The $y$ value to be used in the current event is determined. If
possible, this is the value given by you, but additional
constraints exist from the validity of the parameterizations
($y \geq 0.001$ for GKS, $0.01 \leq y \leq 0.05$ for ERT(Zhu))
and an extra (user-modifiable) requirement of a minimum absolute
invariant mass between jets (which translates into varying $y$ cuts
due to the effects of initial-state QED radiation).
\item
The $\alphas$ value is calculated.
\item
For the $y$ and $\alphas$ values given, the relative
two/three/four-jet composition is determined. This is achieved by
using the parameterized functions of $y$ for 3- and 4-jet rates,
multiplied by the relevant number of factors of $\alphas$.
In ERT(Zhu), where the second-order 3-jet rate is available
only at a few $y$ values, intermediate results are obtained by linear
interpolation in the ratio of second-order to first-order
3-jet rates. The 3-jet and 4-jet rates are normalized to
the analytically known second-order total event rate, i.e.\ divided
by $R_{\mrm{QCD}}$ of eq.~(\ref{ee:RQCD}). Finally, the 2-jet rate is
obtained by conservation of total probability.
\item
If the combination of $y$ and $\alphas$ values is such that the total
3- plus 4-jet fraction is larger than unity, i.e.\ the remainder
2-jet fraction negative, the $y$-cut value is raised (for that event),
and the process is started over at point 3.
\item
The choice is made between generating a 2-, 3- or 4-jet event,
according to the relative probabilities.
\item
For the generation of 4-jets, it is first necessary to make a choice
between $\q \qbar \g \g$ and $\q \qbar \q' \qbar'$ events, according to
the relative (parameterized) total cross sections. A phase-space point
is then selected, and the differential cross section at this point is
evaluated and compared with a parameterized maximum weight. If the
phase-space point is rejected, a new one is selected, until an
acceptable 4-jet event is found.
\item
For 3-jets, a phase-space point is first chosen according to the
first-order cross section. For this point, the weight
\begin{equation}
W(x_1,x_2;y) = 1 + \frac{\alphas}{\pi} R(x_1,x_2;y)
\label{ee:WTJS}
\end{equation}
is evaluated. Here $R(x_1,x_2;y)$ is analytically given for GKS
\cite{Gut84}, while it is approximated by the parameterization
$F(X,Y;y)$ of eq.~(\ref{ee:Zhupar}) for ERT(Zhu). Again, linear
interpolation of $F(X,Y;y)$ has to be applied for intermediate $y$
values. The weight $W$ is compared with a maximum weight
\begin{equation}
W_{\mmax}(y) = 1 + \frac{\alphas}{\pi} R_{\mmax}(y) ~,
\end{equation}
which has been numerically determined beforehand and suitably
parameterized. If the phase-space point is rejected, a
new point is generated, etc.
\item
Massive matrix elements are not implemented for second-order
QCD (but are in the first-order option). However, if a
3- or 4-jet event determined above falls outside
the phase-space region allowed for massive quarks, the event is
rejected and reassigned to be a 2-jet event. (The way the
$y_{ij}$ and $y_{ijk}$ variables of 4-jet events should be
interpreted for massive quarks is not even unique, so some latitude
has been taken here to provide a reasonable continuity from
3-jet events.) This procedure is known not to give the expected full
mass suppression, but is a reasonable first approximation.
\item
Finally, if the event is classified as a 2-jet event, either
because it was initially so assigned, or because it failed the
massive phase-space cuts for 3- and 4-jets, the
generation of 2-jets is trivial.
\end{Enumerate}
\subsubsection{Optimized perturbation theory}
\label{sss:optimizedpt}
Theoretically, it turns out that the second-order corrections to the
3-jet rate are large. It is therefore not unreasonable to expect
large third-order corrections to the 4-jet rate. Indeed, the
experimental 4-jet rate is much larger than second order predicts
(when fragmentation effects have been included),
if $\alphas$ is determined based on the 3-jet rate
\cite{Sjo84a,JAD88}.
The only consistent way to resolve this issue is to go ahead and
calculate the full next order. This is a tough task, however, so
people have looked at possible shortcuts.
For example, one can try to minimize the higher-order contributions
by a suitable choice of the renormalization scale \cite{Ste81} ---
`optimized perturbation theory'. This
is equivalent to a different choice for the $Q^2$ scale in
$\alphas$, a scale which is not unambiguous anyway. Indeed
the standard value $Q^2 = s = E_{\mrm{cm}}^2$ is larger than the
natural physical scale of gluon emission in events, given that most
gluons are fairly soft. One could therefore pick another scale,
$Q^2 = f s$, with $f < 1$. The
${\cal O}(\alphas)$ 3-jet rate would be increased by
such a scale change, and so would the number of 4-jet
events, including those which collapse into 3-jet ones. The loop
corrections depend on the $Q^2$ scale, however,
and compensate the changes above by giving a larger negative
contribution to the 3-jet rate.
The possibility of picking an optimized scale $f$ is implemented
as follows \cite{Sjo89}. Assume that the differential 3-jet
rate at scale $Q^2 = s$ is given by the expression
\begin{equation}
R_3 = r_1 \alphas + r_2 \alphas^2 ~,
\end{equation}
where $R_3$, $r_1$ and $r_2$ are functions of the kinematical
variables $x_1$ and $x_2$ and the $y$ cut, as implied by the
second-order formulae above, see e.g.\ eq.~(\ref{ee:Zhupar}).
When the coupling is chosen at a different scale, $Q'^2 = f s$,
the 3-jet rate has to be changed to
\begin{equation}
R_3' = r_1' \alphas' + r_2' \alphas'^2 ~,
\end{equation}
where $r_1' = r_1$,
\begin{equation}
r_2' = r_2 + r_1 \frac{33-2n_f}{12\pi} \ln f ~,
\label{ee:r2optim}
\end{equation}
and $\alphas' = \alphas(fs)$.
Since we only have the Born term for 4-jets, here the effects of a
scale change come only from the change in the coupling constant.
Finally, the 2-jet cross section can still be calculated from the
difference between the total cross section and the 3- and 4-jet
cross sections.
If an optimized scale is used in the program, the default value is
$f=0.002$, which is favoured by the studies in ref. \cite{Bet89}. (In
fact, it is also possible to use a correspondingly optimized
$R_{\mrm{QCD}}$ factor, eq.~(\ref{ee:RQCD}), but then the
corresponding $f$ is chosen independently and much closer to unity.)
The success of describing the jet rates should not hide the fact that
one is dabbling in (educated, hopefully) guesswork, and that any
conclusions based on this method have to be taken with a pinch of
salt.
One special problem associated with the use of optimized perturbation
theory is that the differential 3-jet rate may become negative
over large regions of the $(x_1, x_2)$ phase space. This problem
already exists, at least in principle, even for a scale $f = 1$,
since $r_2$ is not guaranteed to be positive definite. Indeed,
depending on the choice of $y$ cut, $\alphas$ value, and recombination
scheme, one may observe a small region of negative differential
3-jet rate for the full second-order expression. This region
is centred around $\q \qbar \g$ configurations, where the $\q$ and
$\qbar$ are close together in one hemisphere and the $\g$ is alone in
the other, i.e.\ $x_1 \approx x_2 \approx 1/2$. It is well understood
why second-order corrections should be negative in this region
\cite{Dok89}: the $\q$ and $\qbar$ of a $\q \qbar \g$ state are in a
relative colour octet state, and thus the colour force between them is
repulsive, which translates into a negative second-order term.
However, as $f$ is decreased below unity, $r_2'$ receives a negative
contribution from the $\ln f$ term, and the region of negative
differential cross section has a tendency to become larger, also
after taking into account related changes in $\alphas$. In an
event-generator framework, where all events are supposed to come
with unit
weight, it is clearly not possible to simulate negative cross sections.
What happens in the program is therefore that no 3-jet events at
all are generated in the regions of negative differential cross section,
and that the 3-jet rate in regions of positive cross sections is
reduced by a constant factor, chosen so that the total number of
3-jet events comes out as it should. This is a consequence of the
way the program works, where it is first decided what kind of event to
generate, based on integrated 3-jet rates in which positive and
negative contributions are added up with sign, and only thereafter
the kinematics is chosen.
Based on our physics understanding of the origin of this negative
cross section, the approach adopted is as sensible as any, at least
to that order in perturbation theory (what one might strive for is a
properly exponentiated description of the relevant region). It can
give rise to funny results for low $f$ values, however, as observed
by OPAL \cite{OPA92} for the energy--energy correlation asymmetry.
\subsubsection{Angular orientation}
While pure $\gamma$ exchange gives a simple $1 + \cos^2\theta$
distribution for the $\q$ (and $\qbar$) direction in $\q \qbar$ events,
$\Z^0$ exchange and $\gammaZ$ interference results in a
forward--backward asymmetry. If one introduces
\begin{eqnarray}
h'_{\f}(s) & = & 2 e_{\e} \left\{ a_{\e}
(1 - P_{\mrm{L}}^+ P_{\mrm{L}}^-) -
v_{\e} (P_{\mrm{L}}^- - P_{\mrm{L}}^+) \right\} \,
\Re\mrm{e}\chi(s) e_{\f} a_{\f}
\nonumber \\
& & + \, \left\{ 2 v_{\e} a_{\e} (1 - P_{\mrm{L}}^+ P_{\mrm{L}}^-) -
(v_{\e}^2 + a_{\e}^2) (P_{\mrm{L}}^- - P_{\mrm{L}}^+) \right\} \,
|\chi(s)|^2 \, v_{\f} a_{\f} ~,
\end{eqnarray}
then the angular distribution of the quark is given by
\begin{equation}
\frac{\d \sigma}{\d (\cos\theta_{\f})} \propto
h_{\f}(s)(1 + \cos^2\theta_{\f}) + 2 h'_{\f}(s) \cos\theta_{\f} ~.
\end{equation}
The angular orientation of a 3- or 4-jet event may be described
in terms of three angles $\chi$, $\theta$ and $\varphi$; for 2-jet
events only $\theta$ and $\varphi$ are necessary. From a standard
orientation, with the $\q$ along the $+z$ axis and the $\qbar$ in the
$xz$ plane with $p_x > 0$, an arbitrary orientation may be reached by
the rotations $+\chi$ in azimuthal angle, $+\theta$ in polar angle,
and $+\varphi$ in azimuthal angle, in that order. Differential
cross sections,
including QFD effects and arbitrary beam polarizations have been given
for 2- and 3-jet events in refs. \cite{Ols80,Sch80}. We use the
formalism of ref. \cite{Ols80}, with translation from their
terminology according to $\chi \to \pi - \chi$ and
$\varphi^- \to - (\varphi + \pi/2)$. The resulting formulae are
tedious, but
straightforward to apply, once the internal jet configuration has been
chosen. 4-jet events are approximated by 3-jet ones, by joining
the two gluons of a $\q \qbar \g \g$ event and the $\q'$ and $\qbar'$
of a $\q \qbar \q' \qbar'$ event into one effective jet. This means
that some angular asymmetries are neglected \cite{Ali80a}, but that weak
effects are automatically included. It is assumed that the second-order
3-jet events have the same angular orientation as the first-order
ones, some studies on this issue may be found in \cite{Kor85}. Further,
the formulae normally refer to the massless case; only for the QED
2- and 3-jet cases are mass corrections available.
The main effect of the angular distribution of multijet events
is to smear the lowest-order result, i.e.\ to reduce any anisotropies
present in 2-jet systems. In the parton-shower option of the program,
only the initial $\q \qbar$ axis is determined. The subsequent shower
evolution then {\it de facto} leads to a smearing of the jet axis,
although not necessarily in full agreement with the expectations
from multijet matrix-element treatments.
\subsubsection{Initial-state radiation}
Initial-state photon radiation has been included using the formalism of
ref. \cite{Ber82}. Here each event contains either no photon or one,
i.e.\ it is a first-order non-exponentiated description.
The main formula for the hard radiative photon
cross section is
\begin{equation}
\frac{\d \sigma}{\d x_{\gamma}} = \frac{\alphaem}{\pi} \,
\left( \ln\frac{s}{m_{\e}^2} -1 \right) \,
\frac{1 + (1-x_{\gamma})^2}{x_{\gamma}} \, \sigma_0 (\hat{s}) ~,
\end{equation}
where $x_{\gamma}$ is the photon energy fraction of the beam energy,
$\hat{s} = (1-x_{\gamma}) s$ is the squared reduced hadronic c.m.\
energy, and $\sigma_0$ is the ordinary annihilation cross section at
the reduced energy. In particular, the selection of jet flavours
should be done according to expectations at the reduced
energy. The cross section is divergent both for $x_{\gamma} \to 1$ and
$x_{\gamma} \to 0$. The former is related to the fact that
$\sigma_0$ has a $1/\hat{s}$ singularity (the real photon pole) for
$\hat{s} \to 0$. An upper cut on $x_{\gamma}$ can here be chosen
to fit the
experimental setup. The latter is a soft photon singularity, which is
to be compensated in the no-radiation cross section. A requirement
$x_{\gamma} > 0.01$ has therefore been chosen so that the hard-photon
fraction is smaller than unity. In the total cross section, effects
from photons with $x_{\gamma} < 0.01$ are taken into account, together
with vertex and vacuum polarization corrections (hadronic vacuum
polarizations using a simple parameterization of the more complicated
formulae of ref. \cite{Ber82}).
The hard photon spectrum can be integrated analytically, for the
full $\gammaZ$ structure including interference terms, provided that
no new flavour thresholds are crossed and that the $R_{\mrm{QCD}}$
term in the cross section can be approximated by a constant over the
range of allowed $\hat{s}$ values. In fact, threshold effects can be
taken into account by standard rejection techniques, at the price of
not obtaining the exact cross section analytically, but only by an
effective Monte Carlo integration taking place in parallel with the
ordinary event generation. In addition to $x_{\gamma}$, the polar
angle $\theta_{\gamma}$ and azimuthal angle $\varphi_{\gamma}$ of
the photons are also to be chosen. Further, for the orientation
of the hadronic system, a choice has to be made whether the photon is
to be considered as having been radiated from the $\e^+$ or from the
$\e^-$.
Final-state photon radiation, as well as interference between initial-
and final-state radiation, has been left out of this treatment. The
formulae for $\ee \to \mu^+ \mu^-$ cannot be simply taken over for
the case of outgoing quarks, since the quarks as such only live for
a short while before turning into hadrons. Another simplification in
our treatment is that effects of incoming polarized $\e^{\pm}$ beams
have been completely neglected, i.e.\ neither the effective shift in
azimuthal distribution of photons nor the reduction in polarization
is included. The polarization parameters of the program are to be
thought of as the effective polarization surviving after
initial-state radiation.
\subsubsection{Alternative matrix elements}
The program contains two sets of `toy model' matrix elements, one
for an Abelian vector gluon model and one for a scalar gluon model.
Clearly both of these alternatives are already excluded by data,
and are anyway not viable alternatives for a consistent theory of
strong interactions. They are therefore included more as references
to show how well the characteristic features of QCD can be measured
experimentally.
Second-order matrix elements are available for the Abelian vector
gluon model. These are easily obtained from the standard QCD matrix
elements by a substitution of the Casimir group factors:
$C_F = 4/3 \to 1$, $N_C = 3 \to 0$, and $T_R = n_{\f}/2 \to 3 n_{\f}$.
First-order matrix elements contain only $C_F$; therefore the
standard first-order QCD results may be recovered by a rescaling
of $\alphas$ by a factor $4/3$. In second order the change of $N_C$
to 0 means that $\g \to \g\g$ couplings are absent from the Abelian
model, while the change of $T_R$ corresponds to an enhancement of the
$\g \to \q'\qbar'$ coupling, i.e.\ to an enhancement of the
$\q\qbar\q'\qbar'$ 4-jet event rate.
The second-order corrections to the 3-jet rate
turn out to be strongly negative --- if
$\alphas$ is fitted to get about the right rate of 4-jet events,
the predicted differential 3-jet rate is negative almost
everywhere in the $(x_1, x_2)$ plane. Whether this unphysical
behaviour would be saved by higher orders is unclear. It has been
pointed out that the rate can be made positive by a suitable choice of
scale, since $\alphas$ runs in opposite directions in an Abelian
model and in QCD \cite{Bet89}. This may be seen directly from
eq.~(\ref{ee:r2optim}), where the term $33 = 11 N_C$ is absent in the
Abelian model, and therefore the scale-dependent term changes sign.
In the program, optimized scales have not been implemented for this
toy model. Therefore the alternatives provided for you are either
to generate only 4-jet events, or to neglect second-order
corrections to the 3-jet rate, or to have the total 3-jet
rate set vanishing (so that only 2- and 4-jet events are
generated). Normally we would expect the former to be the one of most
interest, since it is in angular (and flavour) distributions of 4-jet
events that the structure of QCD can be tested.
Also note that the `correct' running of $\alphas$ is not included;
you are expected to use the option where $\alphas$ is just given as
a constant number.
The scalar gluon model is even more excluded than the Abelian vector
one, since differences appear already in the 3-jet matrix
element \cite{Lae80}:
\begin{equation}
\frac{\d \sigma}{\d x_1 \, \d x_2} \propto \frac{x_3^2}{(1-x_1)(1-x_2)}
\end{equation}
when only $\gamma$ exchange is included. The axial part of the $\Z^0$
gives a slightly different shape; this is included in the program but
does not make much difference. The angular orientation does include
the full $\gammaZ$ interference \cite{Lae80}, but the main interest is
in the 3-jet topology as such \cite{Ell79}. No higher-order
corrections are included. It is recommended to use the option of a
fixed $\alphas$ also here, since the correct running is not available.
\subsection{Decays of Onia Resonances}
\label{ss:oniadecays}
Many different possibilities are open for the decay of heavy
$J^{PC} = 1^{--}$ onia resonances. Of special interest are
the decays into three gluons or two gluons plus a photon, since these
offer unique possibilities to study a `pure sample' of gluon jets.
A routine for this purpose is included in the program. It was written
at a time where the expectations were to find toponium at PETRA
energies. Given the large value of the top mass, weak decays
dominate, to the extent that the top quark decays weakly even
before a bound toponium state is formed, and thus the routine will be
of no use for top. The charm system, on the other hand, is far too low
in mass for a jet language to be of any use. The only application is
therefore likely to be for $\Upsilon$, which unfortunately also is on
the low side in mass.
The matrix element for $\q \qbar \to \g \g \g$ is (in lowest order)
\cite{Kol78}
\begin{equation}
\frac{1}{\sigma_{\g \g \g}}
\frac{\d \sigma_{\g \g \g}}{\d x_1 \, \d x_2} =
\frac{1}{\pi^2 - 9} \left\{ \left( \frac{1-x_1}{x_2 x_3} \right)^2 +
\left( \frac{1-x_2}{x_1 x_3} \right)^2 +
\left( \frac{1-x_3}{x_1 x_2} \right)^2 \right\} ~,
\label{ee:Upsilondec}
\end{equation}
where, as before, $x_i = 2 E_i / E_{\mrm{cm}}$ in the c.m.\ frame of
the event. This is a well-defined expression, without the kind of
singularities encountered in the $\q \qbar \g$ matrix elements.
In principle, no cuts at all would be necessary, but for reasons
of numerical simplicity we implement a $y$ cut as for continuum
jet production, with all events not fulfilling this cut considered
as (effective) $\g \g$ events. For $\g \g \g$ events, each $\g \g$
invariant mass is required to be at least 2 GeV.
Another process is $\q \qbar \to \gamma \g \g$, obtained by replacing
a gluon in $\q \qbar \to \g \g \g$ by a photon. This process has the
same normalized cross section as the one above, if e.g.\ $x_1$ is taken
to refer to the photon. The relative rate is \cite{Kol78}
\begin{equation}
\frac{\sigma_{\gamma \g \g}}{\sigma_{\g \g \g}} =
\frac{36}{5} \, \frac{e_{\q}^2 \, \alphaem}
{\alphas(Q^2)} ~.
\label{ee:UpsilonBR}
\end{equation}
Here $e_{\q}$ is the charge of the heavy quark, and the scale in
$\alphas$ has been chosen as the mass of the onium state. If the
mass of the recoiling $\g \g$ system is lower than some cut-off
(by default 2 GeV), the event is rejected.
In the present implementation the angular orientation of the
$\g \g \g$ and $\gamma \g \g$ events is given for the
$\ee \to \gamma^* \to$ onium case \cite{Kol78} (optionally with
beam polarization effects included), i.e.\ weak effects have not
been included, since they are negligible at around 10~GeV.
It is possible to start a perturbative shower evolution from either
of the two states above. However, for $\Upsilon$ the phase space
for additional evolution is so constrained that not much is to be
gained from that. We therefore do not recommend this possibility.
The shower generation machinery, when starting up from a
$\gamma \g \g$ configuration, is constructed such that the
photon energy is not changed. This means that there is currently no
possibility to use showers to bring the theoretical photon
spectrum in better agreement with the experimental one.
In string fragmentation language, a $\g \g \g$ state corresponds to
a closed string triangle with the three gluons at the corners. As
the partons move apart from a common origin, the string triangle
expands. Since the photon does not take part in the fragmentation,
the $\gamma \g \g$ state corresponds to a double string running
between the two gluons.
\subsection{Routines and Common-Block Variables}
\label{ss:eeroutines}
\subsubsection{$\ee$ continuum event generation}
The only routine a normal user will call to generate $\ee$ continuum
events is \ttt{PYEEVT}. The other routines listed below, as well as
\ttt{PYSHOW} (see section \ref{ss:showrout}), are called by
\ttt{PYEEVT}.
\drawbox{CALL PYEEVT(KFL,ECM)}\label{p:PYEEVT}
\begin{entry}
\itemc{Purpose:} to generate a complete event
$\ee \to \gammaZ \to \q\qbar \to$ parton shower $\to$ hadrons
according to QFD and QCD cross sections. As an alternative to
parton showers, second-order matrix elements are available for
$\q\qbar + \q\qbar\g + \q\qbar\g\g + \q\qbar\q'\qbar'$ production.
\iteme{KFL :} flavour of events generated.
\begin{subentry}
\iteme{= 0 :} mixture of all allowed flavours according to relevant
probabilities.
\iteme{= 1 - 8 :} primary quarks are only of the specified flavour
\ttt{KFL}.
\end{subentry}
\iteme{ECM :} total c.m.\ energy of system.
\itemc{Remark:} Each call generates one event, which is independent
of preceding ones, with one exception, as follows. If radiative
corrections are included, the shape of the hard photon spectrum is
recalculated only with each \ttt{PYXTEE} call, which normally is done
only if \ttt{KFL}, \ttt{ECM} or \ttt{MSTJ(102)} is changed. A change
of e.g.\ the $\Z^0$ mass in mid-run has to be followed either by a user
call to \ttt{PYXTEE} or by an internal call forced e.g.\ by putting
\ttt{MSTJ(116) = 3}.
\end{entry}
\boxsep
\begin{entry}
\iteme{SUBROUTINE PYXTEE(KFL,ECM,XTOT) :}\label{p:PYXTEE}
to calculate the total hadronic cross section,
including quark thresholds, weak, beam polarization, and QCD effects
and radiative corrections. In the process, variables necessary
for the treatment of hard photon radiation are calculated and
stored.
\begin{subentry}
\iteme{KFL, ECM :} as for \ttt{PYEEVT}.
\iteme{XTOT :} the calculated total cross section in nb.
\end{subentry}
\iteme{SUBROUTINE PYRADK(ECM,MK,PAK,THEK,PHIK,ALPK) :}\label{p:PYRADK}
to describe initial-state hard $\gamma$ radiation.
\iteme{SUBROUTINE PYXKFL(KFL,ECM,ECMC,KFLC) :}\label{p:PYXKFL}
to generate the primary quark flavour in case this
is not specified by you.
\iteme{SUBROUTINE PYXJET(ECM,NJET,CUT) :}\label{p:PYXJET}
to determine the number of jets (2, 3 or 4) to be
generated within the kinematically allowed region (characterized by
\ttt{CUT} $= y_{\mrm{cut}}$) in the matrix-element approach; to be
chosen such that all probabilities are between 0 and 1.
\iteme{SUBROUTINE PYX3JT(NJET,CUT,KFL,ECM,X1,X2) :}\label{p:PYX3JT}
to generate the internal momentum variables of a
3-jet event, $\q\qbar\g$, according to first- or second-order
QCD matrix elements.
\iteme{SUBROUTINE PYX4JT(NJET,CUT,KFL,ECM,KFLN,X1,X2,X4,X12,X14) :}%
\label{p:PYX4JT}
to generate the internal momentum variables for a
4-jet event, $\q\qbar\g\g$ or $\q\qbar\q'\qbar'$, according to
second-order QCD matrix elements.
\iteme{SUBROUTINE PYXDIF(NC,NJET,KFL,ECM,CHI,THE,PHI) :}\label{p:PYXDIF}
to describe the angular orientation of the jets.
In first-order QCD the complete QED or QFD formulae are used; in
second order 3-jets are assumed to have the same orientation
as in first, and 4-jets are approximated by 3-jets.
\end{entry}
\subsubsection{A routine for onium decay}
In \ttt{PYONIA} we have implemented the decays of heavy onia
resonances into three gluons or two gluons plus a photon, which are
the dominant non-background-like decays of $\Upsilon$.
\drawbox{CALL PYONIA(KFL,ECM)}\label{p:PYONIA}
\begin{entry}
\itemc{Purpose:} to simulate the process
$\ee \to \gamma^* \to 1^{--}$ onium resonance $\to (\g\g\g$ or
$\g\g\gamma) \to$ shower $\to$ hadrons.
\iteme{KFL :} the flavour of the quark giving rise to the resonance.
\begin{subentry}
\iteme{= 0 :} generate $\g\g\g$ events alone.
\iteme{= 1 - 8 :} generate $\g\g\g$ and $\g\g\gamma$ events in mixture
determined by the squared charge of flavour \ttt{KFL}, see
eq.~(\ref{ee:UpsilonBR}). Normally \ttt{KFL =} 5.
\end{subentry}
\iteme{ECM :} total c.m.\ energy of system.
\end{entry}
\subsubsection{Common-block variables}
The status codes and parameters relevant for the $\ee$ routines are
found in the common block \ttt{PYDAT1}. This common block also contains
more general status codes and parameters, described elsewhere.
\drawbox{COMMON/PYDAT1/MSTU(200),PARU(200),MSTJ(200),PARJ(200)}
\begin{entry}
\itemc{Purpose:} to give access to a number of status codes and
parameters regulating the performance of the $\ee$ event generation
routines.
\iteme{MSTJ(101) :}\label{p:MSTJ101} (D = 5) gives the type of QCD
corrections used for continuum events.
\begin{subentry}
\iteme{= 0 :} only $\q\qbar$ events are generated.
\iteme{= 1 :} $\q\qbar + \q\qbar\g$ events are generated according
to first-order QCD.
\iteme{= 2 :} $\q\qbar + \q\qbar\g + \q\qbar\g\g + \q\qbar\q'\qbar'$
events are generated according to second-order QCD.
\iteme{= 3 :} $\q\qbar + \q\qbar\g + \q\qbar\g\g + \q\qbar\q'\qbar'$
events are generated, but without second-order corrections to the
3-jet rate.
\iteme{= 5 :} a parton shower is allowed to develop from an original
$\q\qbar$ pair, see \ttt{MSTJ(38) - MSTJ(50)} for details.
\iteme{= -1 :} only $\q\qbar\g$ events are generated (within same
matrix-element cuts as for \ttt{= 1}). Since the change in flavour
composition from mass cuts or radiative corrections is not
taken into account, this option is not intended for
quantitative studies.
\iteme{= -2 :} only $\q\qbar\g\g$ and $\q\qbar\q'\qbar'$ events are
generated (as for \ttt{= 2}). The same warning as for \ttt{= -1} applies.
\iteme{= -3 :} only $\q\qbar\g\g$ events are generated (as for
\ttt{= 2}). The same warning as for \ttt{= -1} applies.
\iteme{= -4 :} only $\q\qbar\q'\qbar'$ events are generated
(as for \ttt{= 2}). The same warning as for \ttt{= -1} applies.
\itemc{Note 1:} \ttt{MSTJ(101)} is also used in \ttt{PYONIA}, with
\iteme{$\leq$ 4 :} $\g\g\g + \gamma\g\g$ events are generated
according to lowest-order matrix elements.
\iteme{$\geq$ 5 :} a parton shower is allowed to develop from the
original $\g\g\g$ or $\g\g\gamma$ configuration, see
\ttt{MSTJ(38) - MSTJ(50)} for details.
\itemc{Note 2:} the default values of fragmentation parameters have
been chosen to work well with the default parton-shower approach
above. If any of the other options are used, or if the parton
shower is used in non-default mode, it is normally necessary to
retune fragmentation parameters. As an example, we note that
the second-order matrix-element approach (\ttt{MSTJ(101) = 2}) at
PETRA/PEP energies gives a better description when the $a$ and
$b$ parameters of the symmetric fragmentation function are set to
$a =$\ttt{PARJ(41) = 1}, $b =$\ttt{PARJ(42) = 0.7}, and the
width of the transverse momentum distribution to
$\sigma =$\ttt{PARJ(21) = 0.40}.
In principle, one also ought to change the joining parameter
to \ttt{PARJ(33) = PARJ(35) = 1.1} to preserve a flat rapidity
plateau, but if this should be forgotten, it does not make too
much difference. For applications at TRISTAN or LEP, one has to
change the matrix-element approach
parameters even more, to make up for additional soft gluon
effects not covered in this approach.
\end{subentry}
\iteme{MSTJ(102) :} (D = 2) inclusion of weak effects ($\Z^0$ exchange)
for flavour production, angular orientation, cross sections and
initial-state photon radiation in continuum events.
\begin{subentry}
\iteme{= 1 :} QED, i.e.\ no weak effects are included.
\iteme{= 2 :} QFD, i.e.\ including weak effects.
\iteme{= 3 :} as \ttt{= 2}, but at initialization in \ttt{PYXTEE} the
$\Z^0$ width is calculated from $\ssintw$, $\alphaem$ and
$\Z^0$ and quark masses (including bottom and top threshold factors for
\ttt{MSTJ(103)} odd), assuming three full generations, and the
result is stored in \ttt{PARJ(124)}.
\end{subentry}
\iteme{MSTJ(103) :} (D = 7) mass effects in continuum matrix elements,
in the form \ttt{MSTJ(103)} $= M_1 + 2M_2 + 4M_3$, where $M_i = 0$
if no mass effects and $M_i = 1$ if mass effects should be included.
Here;
\begin{subentry}
\iteme{$M_1$ :} threshold factor for new flavour production
according to QFD result;
\iteme{$M_2$ :} gluon emission probability (only applies for
\ttt{|MSTJ(101)|}$\leq 1$, otherwise no mass effects anyhow);
\iteme{$M_3$ :} angular orientation of event (only applies for
\ttt{|MSTJ(101)|}$\leq 1$ and
\ttt{MSTJ(102) = 1}, otherwise no mass effects anyhow).
\end{subentry}
\iteme{MSTJ(104) :} (D = 5) number of allowed flavours, i.e.\ flavours
that can be produced in a continuum event if the energy is enough.
A change to 6 makes top production allowed above the threshold, etc.
Note that in $\q\qbar\q'\qbar'$ events only the first five flavours
are allowed in the secondary pair, produced by a gluon breakup.
\iteme{MSTJ(105) :} (D = 1) fragmentation and decay in \ttt{PYEEVT} and
\ttt{PYONIA} calls.
\begin{subentry}
\iteme{= 0 :} no \ttt{PYEXEC} calls, i.e.\ only matrix-element
and/or parton-shower treatment, and collapse of small jet
systems into one or two particles (in \ttt{PYPREP}).
\iteme{= 1 :} \ttt{PYEXEC} calls are made to generate fragmentation
and decay chain.
\iteme{= -1 :} no \ttt{PYEXEC} calls and no collapse of small jet
systems into one or two particles (in \ttt{PYPREP}).
\end{subentry}
\iteme{MSTJ(106) :} (D = 1) angular orientation in \ttt{PYEEVT} and
\ttt{PYONIA}.
\begin{subentry}
\iteme{= 0 :} standard orientation of events, i.e.\ $\q$ along $+z$ axis
and $\qbar$ along $-z$ axis or in $xz$ plane with $p_x > 0$ for
continuum events, and $\g_1\g_2\g_3$ or $\gamma\g_2\g_3$ in $xz$ plane
with $\g_1$ or $\gamma$ along the $+z$ axis for onium events.
\iteme{= 1 :} random orientation according to matrix elements.
\end{subentry}
\iteme{MSTJ(107) :} (D = 0) radiative corrections to continuum events.
\begin{subentry}
\iteme{= 0 :} no radiative corrections.
\iteme{= 1 :} initial-state radiative corrections (including weak
effects for \ttt{MSTJ(102) =} 2 or 3).
\end{subentry}
\iteme{MSTJ(108) :} (D = 2) calculation of $\alphas$ for matrix-element
alternatives. The \ttt{MSTU(111)} and \ttt{PARU(112)} values are
automatically overwritten in \ttt{PYEEVT} or \ttt{PYONIA} calls
accordingly.
\begin{subentry}
\iteme{= 0 :} fixed $\alphas$ value as given in \ttt{PARU(111)}.
\iteme{= 1 :} first-order formula is always used, with
$\Lambda_{\mrm{QCD}}$ given by \ttt{PARJ(121)}.
\iteme{= 2 :} first- or second-order formula is used, depending on
value of \ttt{MSTJ(101)}, with $\Lambda_{\mrm{QCD}}$ given by
\ttt{PARJ(121)} or \ttt{PARJ(122)}.
\end{subentry}
\iteme{MSTJ(109) :} (D = 0) gives a possibility to switch from QCD
matrix elements to some alternative toy models. Is not relevant for
shower evolution, \ttt{MSTJ(101) = 5}, where one can use
\ttt{MSTJ(49)} instead.
\begin{subentry}
\iteme{= 0 :} standard QCD scenario.
\iteme{= 1 :} a scalar gluon model. Since no second-order corrections
are available in this scenario, one can only use this with
\ttt{MSTJ(101) = 1} or \ttt{-1}. Also note that the event-as-a-whole
angular distribution is for photon exchange only (i.e.\ no weak
effects), and that no higher-order corrections to the total
cross section are included.
\iteme{= 2 :} an Abelian vector gluon theory, with the colour factors
$C_F = 1$ ($= 4/3$ in QCD), $N_C = 0$ ($= 3$ in QCD) and
$T_R = 3 n_f$ ($= n_f/2$ in QCD). If one selects
$\alpha_{\mrm{Abelian}} = (4/3) \alpha_{\mrm{QCD}}$,
the 3-jet cross section will agree with
the QCD one, and differences are to be found only in 4-jets.
The \ttt{MSTJ(109) = 2} option has to be run with
\ttt{MSTJ(110) = 1} and \ttt{MSTJ(111) = 0}; if need be, the latter
variables will be overwritten by the program. \\
{\bf Warning:} second-order corrections give a large negative
contribution to the 3-jet cross section, so large that
the whole scenario is of doubtful use. In order to make the
second-order options work at all, the 3-jet cross section
is here by hand set exactly equal to zero for \ttt{MSTJ(101) = 2}.
It is here probably better to use the option \ttt{MSTJ(101) = 3},
although this is not a consistent procedure either.
\end{subentry}
\iteme{MSTJ(110) :} (D = 2) choice of second-order contributions
to the 3-jet rate.
\begin{subentry}
\iteme{= 1 :} the GKS second-order matrix elements.
\iteme{= 2 :} the Zhu parameterization of the ERT matrix elements,
based on the program of Kunszt and Ali, i.e.\ in historical sequence
ERT/Kunszt/Ali/Zhu. The parameterization is available for
$y =$ 0.01, 0.02, 0.03, 0.04 and 0.05. Values outside this
range are put at the nearest border, while those inside
it are given by a linear interpolation between the
two nearest points. Since this procedure is rather primitive,
one should try to work at one of the values given above.
Note that no Abelian QCD parameterization is available for
this option.
\end{subentry}
\iteme{MSTJ(111) :} (D = 0) use of optimized perturbation theory for
second-order matrix elements (it can also be used for first-order
matrix elements, but here it only corresponds to a trivial
rescaling of the $\alphas$ argument).
\begin{subentry}
\iteme{= 0 :} no optimization procedure; i.e.\ $Q^2 = E_{\mrm{cm}}^2$.
\iteme{= 1 :} an optimized $Q^2$ scale is chosen as
$Q^2 = f E_{\mrm{cm}}^2$, where $f =$\ttt{PARJ(128)} for the total
cross section $R$ factor, while $f =$\ttt{PARJ(129)} for the
3- and 4-jet rates. This $f$ value enters via the
$\alphas$, and also via a term proportional to $\alphas^2 \ln f$.
Some constraints are imposed; thus the optimized `3-jet'
contribution to $R$ is assumed to be positive (for \ttt{PARJ(128)}),
the total 3-jet rate is not allowed to be negative
(for \ttt{PARJ(129)}), etc.
However, there is no guarantee that the differential 3-jet
cross section is not negative (and truncated to 0) somewhere
(this can also happen with $f = 1$, but is then less frequent).
The actually obtained $f$ values are stored in \ttt{PARJ(168)} and
\ttt{PARJ(169)}, respectively.
If an optimized $Q^2$ scale is used, then the $\Lambda_{\mrm{QCD}}$
(and $\alphas$) should also be changed. With the value $f = 0.002$,
it has been shown \cite{Bet89} that a $\Lambda_{\mrm{QCD}} = 0.100$
GeV gives a reasonable agreement; the parameter to be changed is
\ttt{PARJ(122)} for a second-order running $\alphas$. Note that,
since the optimized $Q^2$ scale is sometimes below the charm
threshold, the effective number of flavours used in $\alphas$ may
well be 4 only. If one feels that it is still appropriate to use 5
flavours (one choice might be as good as the other), it is
necessary to put \ttt{MSTU(113) = 5}.
\end{subentry}
\iteme{MSTJ(115) :} (D = 1) documentation of continuum or onium
events, in increasing order of completeness.
\begin{subentry}
\iteme{= 0 :} only the parton shower, the fragmenting partons and the
generated had\-ronic system are stored in the \ttt{PYJETS} common block.
\iteme{= 1 :} also a radiative photon is stored (for continuum events).
\iteme{= 2 :} also the original $\ee$ are stored (with
\ttt{K(I,1) = 21}).
\iteme{= 3 :} also the $\gamma$ or $\gammaZ$ exchanged for continuum
events, the onium state for resonance events is stored (with
\ttt{K(I,1) = 21}).
\end{subentry}
\iteme{MSTJ(116) :} (D = 1) initialization of total cross section and
radiative photon spectrum in \ttt{PYEEVT} calls.
\begin{subentry}
\iteme{= 0 :} never; cannot be used together with radiative
corrections.
\iteme{= 1 :} calculated at first call and then whenever \ttt{KFL}
or \ttt{MSTJ(102)} is changed or \ttt{ECM} is changed by more than
\ttt{PARJ(139)}.
\iteme{= 2 :} calculated at each call.
\iteme{= 3 :} everything is re-initialized in the next call, but
\ttt{MSTJ(116)} is afterwards automatically put \ttt{= 1} for use
in subsequent calls.
\end{subentry}
\iteme{MSTJ(119) :} (I) check on need to re-initialize \ttt{PYXTEE}.
\iteme{MSTJ(120) :} (R) type of continuum event generated with the
matrix-element option (with the shower one, the result is always
\ttt{= 1}).
\begin{subentry}
\iteme{= 1 :} $\q\qbar$.
\iteme{= 2 :} $\q\qbar\g$.
\iteme{= 3 :} $\q\qbar\g\g$ from Abelian (QED-like) graphs in
matrix element.
\iteme{= 4 :} $\q\qbar\g\g$ from non-Abelian (i.e.\ containing
triple-gluon coupling) graphs in matrix element.
\iteme{= 5 :} $\q\qbar\q'\qbar'$.
\end{subentry}
\iteme{MSTJ(121) :} (R) flag set if a negative differential
cross section was encountered in the latest \ttt{PYX3JT} call.
Events are still generated, but maybe not quite according to
the distribution one would like (the rate is set to zero in the
regions of negative cross section, and the differential rate
in the regions of positive cross section is rescaled to give
the `correct' total 3-jet rate).
\boxsep
\iteme{PARJ(121) :}\label{p:PARJ121} (D = 1.0 GeV) $\Lambda$ value
used in first-order
calculation of $\alphas$ in the matrix-element alternative.
\iteme{PARJ(122) :} (D = 0.25 GeV) $\Lambda$ values used in second-order
calculation of $\alphas$ in the matrix-element alternative.
\iteme{PARJ(123) :} (D = 91.187 GeV) mass of $\Z^0$ as used in
propagators for the QFD case.
\iteme{PARJ(124) :} (D = 2.489 GeV) width of $\Z^0$ as used in
propagators for the QFD case. Overwritten at initialization if
\ttt{MSTJ(102) = 3}.
\iteme{PARJ(125) :} (D = 0.01) $y_{\mrm{cut}}$, minimum squared scaled
invariant mass of any two partons in 3- or 4-jet events; the main
user-controlled matrix-element cut. \ttt{PARJ(126)} provides an
additional constraint. For each new event, it is additionally
checked that the total 3- plus 4-jet fraction does not
exceed unity; if so the effective $y$ cut will be dynamically
increased. The actual $y$-cut value is stored in
\ttt{PARJ(150)}, event by event.
\iteme{PARJ(126) :} (D = 2. GeV) minimum invariant mass of any two
partons in 3- or 4-jet events; a cut in addition to the one above,
mainly for the case of a radiative photon lowering the hadronic
c.m.\ energy significantly.
\iteme{PARJ(127) :} (D = 1. GeV) is used as a safety margin for small
colour-singlet jet systems, cf. \ttt{PARJ(32)}, specifically
$\q\qbar'$ masses in $\q\qbar\q'\qbar'$ 4-jet events and $\g\g$ mass
in onium $\gamma\g\g$ events.
\iteme{PARJ(128) :} (D = 0.25) optimized $Q^2$ scale for the QCD $R$
(total rate) factor for the \ttt{MSTJ(111) = 1} option is given by
$Q^2 = f E_{\mrm{cm}}^2$, where $f =$\ttt{PARJ(128)}. For various
reasons the actually used $f$ value may be increased compared with
the nominal one; while \ttt{PARJ(128)} gives the nominal value,
\ttt{PARJ(168)} gives the actual one for the current event.
\iteme{PARJ(129) :} (D = 0.002) optimized $Q^2$ scale for the 3-
and 4-jet rate for the \ttt{MSTJ(111) = 1} option is given by
$Q^2 = f E_{\mrm{cm}}^2$, where $f =$\ttt{PARJ(129)}. For various
reasons the actually used $f$ value may be increased compared with
the nominal one; while \ttt{PARJ(129)} gives the nominal value,
\ttt{PARJ(169)} gives the actual one for the current event. The
default value is in agreement with the studies of Bethke \cite{Bet89}.
\iteme{PARJ(131), PARJ(132) :} (D = 2*0.) longitudinal polarizations
$P_{\mrm{L}}^+$ and $P_{\mrm{L}}^-$ of incoming $\e^+$ and $\e^-$.
\iteme{PARJ(133) :} (D = 0.) transverse polarization
$P_{\mrm{T}} = \sqrt{P_{\mrm{T}}^+ P_{\mrm{T}}^-}$, with
$P_{\mrm{T}}^+$ and $P_{\mrm{T}}^-$ transverse
polarizations of incoming $\e^+$ and $\e^-$.
\iteme{PARJ(134) :} (D = 0.) mean of transverse polarization
directions of incoming $\e^+$ and $\e^-$,
$\Delta \varphi = (\varphi^+ + \varphi^-) /2$, with $\varphi$
the azimuthal angle of polarization, leading to a shift in the
$\varphi$ distribution of jets by $\Delta \varphi$.
\iteme{PARJ(135) :} (D = 0.01) minimum photon energy fraction
(of beam energy) in initial-state radiation; should normally
never be changed (if lowered too much, the fraction of events
containing a radiative photon will exceed unity, leading to
problems).
\iteme{PARJ(136) :} (D = 0.99) maximum photon energy fraction
(of beam energy) in initial-state radiation; may be changed
to reflect actual trigger conditions of a detector (but must
always be larger than \ttt{PARJ(135)}).
\iteme{PARJ(139) :} (D = 0.2 GeV) maximum deviation of $E_{\mrm{cm}}$
from the corresponding value at last \ttt{PYXTEE} call, above which
a new call is made if \ttt{MSTJ(116) = 1}.
\iteme{PARJ(141) :} (R) value of $R$, the ratio of continuum
cross section to the lowest-order muon pair production cross section,
as given in massless QED (i.e.\ three times the sum of active
quark squared charges, possibly modified for polarization).
\iteme{PARJ(142) :} (R) value of $R$ including quark-mass effects
(for \ttt{MSTJ(102) = 1}) and/or weak propagator effects
(for \ttt{MSTJ(102) = 2}).
\iteme{PARJ(143) :} (R) value of $R$ as \ttt{PARJ(142)}, but
including QCD corrections as given by \ttt{MSTJ(101)}.
\iteme{PARJ(144) :} (R) value of $R$ as \ttt{PARJ(143)}, but
additionally including corrections from initial-state photon
radiation (if \ttt{MSTJ(107) = 1}). Since the effects of heavy
flavour thresholds are not simply integrable, the initial value
of \ttt{PARJ(144)} is updated during the
course of the run to improve accuracy.
\iteme{PARJ(145) - PARJ(148) :} (R) absolute cross sections in nb
as for the cases \ttt{PARJ(141) - PARJ(144)} above.
\iteme{PARJ(150) :} (R) current effective matrix element cut-off
$y_{\mrm{cut}}$, as given by \ttt{PARJ(125), PARJ(126)} and the
requirements of having non-negative cross sections for 2-,
3- and 4-jet events. Not used in parton showers.
\iteme{PARJ(151) :} (R) value of c.m.\ energy \ttt{ECM} at last
\ttt{PYXTEE} call.
\iteme{PARJ(152) :} (R) current first-order contribution to the
3-jet fraction; modified by mass effects. Not used in parton
showers.
\iteme{PARJ(153) :} (R) current second-order contribution to the
3-jet fraction; modified by mass effects. Not used in parton
showers.
\iteme{PARJ(154) :} (R) current second-order contribution to the
4-jet fraction; modified by mass effects. Not used in parton
showers.
\iteme{PARJ(155) :} (R) current fraction of 4-jet rate
attributable to $\q\qbar\q'\qbar'$ events rather than $\q\qbar\g\g$
ones; modified by mass effects. Not used in parton showers.
\iteme{PARJ(156) :} (R) has two functions when using second-order
QCD. For a 3-jet event, it gives the ratio of the second-order
to the total 3-jet cross section in the given kinematical
point. For a 4-jet event, it gives the ratio of the
modified 4-jet cross section, obtained when neglecting interference
terms whose colour flow is not well defined, to the full
unmodified one, all evaluated in the given kinematical point.
Not used in parton showers.
\iteme{PARJ(157) - PARJ(159) :} (I) used for cross-section
calculations to include mass threshold effects to radiative
photon cross section. What is stored is basic cross section,
number of events generated and number that passed cuts.
\iteme{PARJ(160) :} (R) nominal fraction of events that should
contain a radiative photon.
\iteme{PARJ(161) - PARJ(164) :} (I) give shape of radiative photon
spectrum including weak effects.
\iteme{PARJ(168) :} (R) actual $f$ value of current event in
optimized perturbation theory for $R$; see \ttt{MSTJ(111)} and
\ttt{PARJ(128)}.
\iteme{PARJ(169) :} (R) actual $f$ value of current event in
optimized perturbation theory for 3- and 4-jet rate;
see \ttt{MSTJ(111)} and \ttt{PARJ(129)}.
\iteme{PARJ(171) :} (R) fraction of cross section corresponding
to the axial coupling of quark pair to the intermediate $\gammaZ$
state; needed for the Abelian gluon model 3-jet matrix
element.
\end{entry}
\subsection{Examples}
An ordinary $\ee$ annihilation event in the continuum, at a
c.m.\ energy of 91 GeV, may be generated with
\begin{verbatim}
CALL PYEEVT(0,91D0)
\end{verbatim}
In this case a $\q\qbar$ event is generated, including weak effects,
followed by parton-shower evolution and fragmentation/decay treatment.
Before a call to \ttt{PYEEVT}, however, a number of default values
may be changed, e.g.\ \ttt{MSTJ(101) = 2} to use second-order QCD
matrix elements, giving a mixture of $\q\qbar$, $\q\qbar\g$,
$\q\qbar\g\g$, and $\q\qbar\q'\qbar'$ events, \ttt{MSTJ(102) = 1} to
have QED only, \ttt{MSTJ(104) = 6} to allow $\t\tbar$ production
as well, \ttt{MSTJ(107) = 1} to include initial-state photon radiation
(including a treatment of the $\Z^0$ pole), \ttt{PARJ(123) = 92.0} to
change the $\Z^0$ mass, \ttt{PARJ(81) = 0.3} to change the
parton-shower $\Lambda$ value, or \ttt{PARJ(82) = 1.5} to change the
parton-shower cut-off. If initial-state photon radiation is used, some
restrictions apply to how one can alternate the generation of
events at different energies or with different $\Z^0$ mass, etc.
These restrictions are not there for
efficiency reasons (the extra time for recalculating the extra
constants every time is small), but because it ties in with the
cross-section calculations (see \ttt{PARJ(144)}).
Most parameters can be changed independently of each other. However,
if just one or a few parameters/switches are changed, one should not
be surprised to find a rather bad agreement with the data, like e.g.
a too low or high average hadron multiplicity. It is therefore usually
necessary to retune one parameter related to the perturbative QCD
description, like $\alphas$ or $\Lambda$, one of the two parameters
$a$ and $b$ of the Lund symmetric fragmentation function (since they
are so strongly correlated, it is often not necessary to retune both
of them), and the average fragmentation transverse momentum --- see
Note~2 of the \ttt{MSTJ(101)} description for an example. For very
detailed studies it may be necessary to retune even more parameters.
The three-gluon and gluon--gluon--photon decays of $\Upsilon$ may be
simulated by a call
\begin{verbatim}
CALL PYONIA(5,9.46D0)
\end{verbatim}
A typical program for analysis of $\ee$ annihilation events at
200 GeV might look something like
\begin{verbatim}
IMPLICIT DOUBLE PRECISION(A-H, O-Z)
IMPLICIT INTEGER(I-N)
INTEGER PYK,PYCHGE,PYCOMP
COMMON/PYJETS/N,NPAD,K(4000,5),P(4000,5),V(4000,5)
COMMON/PYDAT1/MSTU(200),PARU(200),MSTJ(200),PARJ(200)
COMMON/PYDAT2/KCHG(500,4),PMAS(500,4),PARF(2000),VCKM(4,4)
COMMON/PYDAT3/MDCY(500,3),MDME(8000,2),BRAT(8000),KFDP(8000,5)
MDCY(PYCOMP(111),1)=0 ! put pi0 stable
MSTJ(107)=1 ! include initial-state radiation
PARU(41)=1D0 ! use linear sphericity
..... ! other desired changes
CALL PYTABU(10) ! initialize analysis statistics
DO 100 IEV=1,1000 ! loop over events
CALL PYEEVT(0,200D0) ! generate new event
IF(IEV.EQ.1) CALL PYLIST(2) ! list first event
CALL PYTABU(11) ! save particle composition
! statistics
CALL PYEDIT(2) ! remove decayed particles
CALL PYSPHE(SPH,APL) ! linear sphericity analysis
IF(SPH.LT.0D0) GOTO 100 ! too few particles in event for
! PYSPHE to work on it (unusual)
CALL PYEDIT(31) ! orient event along axes above
IF(IEV.EQ.1) CALL PYLIST(2) ! list first treated event
..... ! fill analysis statistics
CALL PYTHRU(THR,OBL) ! now do thrust analysis
..... ! more analysis statistics
100 CONTINUE !
CALL PYTABU(12) ! print particle composition
! statistics
..... ! print analysis statistics
END
\end{verbatim}
\clearpage
\section{Process Generation}
\label{s:PYTprocgen}
Much can be said about the processes in {\Py} and
the way they are generated. Therefore the material has been split
into three sections. In the current one the philo\-sophy underlying
the event generation scheme is presented. Here we provide a
generic description, where some special cases are swept under the
carpet. In the next section, the existing processes are enumerated,
with some comments about applications and limitations. Finally, in
the third section the generation routines and common-block switches
are described.
The section starts with a survey of parton distributions, followed
by a detailed description of the simple $2 \to 2$ and $2 \to 1$
hard subprocess generation schemes, including pairs of resonances.
This is followed by a few comments on more complicated configurations,
and on nonperturbative processes.
\subsection{Parton Distributions}
\label{ss:structfun}
The parton distribution function $f_i^a(x,Q^2)$ parameterizes the
probability to find a parton $i$ with a fraction $x$ of the beam energy
when the beam particle $a$ is probed by a hard scattering at virtuality
scale $Q^2$. Usually the momentum-weighted combination $x f_i^a(x,Q^2)$
is used, for which the normalization condition
$\sum_i \int_0^1 dx \, x f_i^a(x,Q^2) \equiv 1$ normally applies.
The $Q^2$ dependence of parton distributions is perturbatively
calculable, see section \ref{sss:initshowstruc}.
The parton distributions in {\Py} come in many shapes, as shown in the
following.
\subsubsection{Baryons}
For protons, many sets exist on the market. These are obtained by fits
to experimental data, constrained so that the $Q^2$ dependence is in
accordance with the standard QCD evolution equations. The current default
in {\Py} is CTEQ 5L \cite{Lai00}, a leading-order fit.
Several other sets are found in {\Py}. The complete list is:
\begin{Itemize}
\item EHLQ sets 1 and 2 \cite{Eic84};
\item DO sets 1 and 2 \cite{Duk82};
\item the GRV 92L (updated version) fit \cite{Glu92};
\item the CTEQ 3L, CTEQ 3M and CTEQ 3D fits \cite{Lai95};
\item the GRV 94L, GRV 94M and GRV 94D fits \cite{Glu95}; and
\item the CTEQ 5L and CTEQ 5M1 fits \cite{Lai00}.
\end{Itemize}
Of these, EHLQ, DO, GRV 92L, CTEQ 3L, GRV94L and CTEQ5L are leading-order
parton distributions, while CTEQ 3D and GRV94D are in the
next-to-leading-order DIS scheme and the rest
in the next-to-leading order $\br{\mrm{MS}}$ scheme. The EHLQ and DO
sets are by now rather old, and are kept mainly for backwards
compatibility. Since only Born-level matrix elements
are included in the program, there is no particular reason to use
higher-order parton distributions --- the resulting combination is
anyway only good to leading-order accuracy. (Some higher-order
corrections are effectively included by the parton-shower
treatment, but there is no exact match.)
There is a steady flow of new parton-distribution sets on the market.
To keep track of all of them is a major work on its own. Therefore
{\Py} contains an interface to an external library of parton
distribution functions, \tsc{Pdflib} \cite{Plo93}. This is an
encyclopedic collection of almost all proton, pion and photon parton
distributions proposed from the late 70's to the late 90's.
Three dummy routines come with the {\Py} package, so as to avoid
problems with unresolved external references if \tsc{Pdflib} is not
linked. One should also note that {\Py} does not check the results,
but assumes that sensible answers will be returned, also outside the
nominal $(x, Q^2)$ range of a set. Only the sets that come with
{\Py} have been suitably modified to provide reasonable answers
outside their nominal domain of validity.
\tsc{Pdflib} has been frozen in recent years. Instead a new project
has taken over the same r\^ole, \tsc{LHAPDF}, the Les Houches Accord
PDF interface \cite{Gie02}, containing all new sets of the last five
years or so. While \tsc{LHAPDF} has a native input/output format
different from the \tsc{Pdflib} one, the \tsc{LHAGLUE} subpackage
allows \tsc{LHAPDF} to be called in exactly the same way as
\tsc{Pdflib} is. Therefore the {\Py} external-PFD-interface
works for \tsc{Pdflib} and \tsc{LHAPDF} alike.
{}From the proton parton distributions, those of the neutron are obtained
by isospin conjugation, i.e.\ $f_{\u}^{\n} = f_{\d}^{\p}$ and
$f_{\d}^{\n} = f_{\u}^{\p}$.
The program does allow for incoming beams of a number of hyperons:
$\Lambda^0$, $\Sigma^{-,0,+}$, $\Xi^{-,0}$ and $\Omega^-$. Here
one has essentially no experimental information. One could imagine
to construct models in which valence $\s$ quarks are found at larger
average $x$ values than valence $\u$ and $\d$ ones, because of the
larger $\s$-quark mass. However, hyperon beams is a little-used part
of the program, included only for a few specific studies. Therefore
a simple approach has been taken, in which an average valence
quark distribution is constructed as
$f_{\mrm{val}} = (f_{\u,\mrm{val}}^{\p} + f_{\d,\mrm{val}}^{\p})/3$,
according to
which each valence quark in a hyperon is assumed to be distributed.
Sea-quark and gluon distributions are taken as in the proton.
Any proton parton distribution set may be used with this procedure.
\subsubsection{Mesons and photons}
Data on meson parton distributions are scarce, so only very few sets
have been constructed, and only for the $\pi^{\pm}$. {\Py} contains
the Owens set 1 and 2 parton distributions \cite{Owe84}, which for a
long time were essentially the only sets on the market, and the
more recent dynamically generated GRV LO (updated version)
\cite{Glu92a}. The latter one is the default in {\Py}. Further sets
are found in \tsc{Pdflib} and \tsc{LHAPDF} and can therefore be used
by {\Py}, just as described above for protons.
Like the proton was used as a template for simple hyperon sets,
so also the pion is used to derive a crude ansatz for
$\K^{\pm}/\K_{\mrm{S}}^0/\K_{\mrm{L}}^0$. The procedure is the
same, except that now $f_{\mrm{val}} =
(f_{\u,\mrm{val}}^{\pi^+} + f_{\dbar,\mrm{val}}^{\pi^+})/2$.
Sets of photon parton distributions have been obtained as for
hadrons; an additional complication comes from the necessity to
handle the matching of the vector meson dominance (VMD) and the
perturbative pieces in a consistent manner. New sets have been
produced where this division is explicit and therefore especially
well suited for applications to event generation\cite{Sch95}.
The Schuler and Sj\"ostand set 1D is the default. Although the
vector-meson philosophy is at the base, the details of the fits do
not rely on pion data, but only on $F_2^{\gamma}$ data. Here
follows a brief summary of relevant details.
Real photons obey a set of inhomogeneous evolution equations, where the
inhomogeneous term is induced by $\gamma \to \q\qbar$ branchings.
The solution can be written as the sum of two terms,
\begin{equation}
f_a^{\gamma}(x,Q^2) = f_a^{\gamma,\mrm{NP}}(x,Q^2;Q_0^2)
+ f_a^{\gamma,\mrm{PT}}(x,Q^2;Q_0^2) ~,
\end{equation}
where the former term is a solution to the homogeneous evolution
with a (nonperturbative) input at $Q=Q_0$ and the latter is a
solution to the full inhomogeneous equation with boundary condition
$f_a^{\gamma,\mrm{PT}}(x,Q_0^2;Q_0^2) \equiv 0$. One possible
physics interpretation is to let $f_a^{\gamma,\mrm{NP}}$ correspond
to $\gamma \leftrightarrow V$ fluctuations, where
$V = \rho^0, \omega, \phi, \ldots$ is a set of vector mesons,
and let $f_a^{\gamma,\mrm{PT}}$ correspond to perturbative (`anomalous')
$\gamma \leftrightarrow \q\qbar$ fluctuations. The discrete spectrum
of vector mesons can be combined with the continuous (in virtuality
$k^2$) spectrum of $\q\qbar$ fluctuations, to give
\begin{equation}
f_a^{\gamma}(x,Q^2) =
\sum_V \frac{4\pi\alphaem}{f_V^2} f_a^{\gamma,V}(x,Q^2)
+ \frac{\alphaem}{2\pi} \, \sum_{\q} 2 e_{\q}^2 \,
\int_{Q_0^2}^{Q^2} \frac{{\d} k^2}{k^2} \,
f_a^{\gamma,\q\qbar}(x,Q^2;k^2) ~,
\label{eq:decomprealga}
\end{equation}
where each component $f^{\gamma,V}$ and $f^{\gamma,\q\qbar}$ obeys a
unit momentum sum rule.
In sets 1 the $Q_0$ scale is picked at a low value, 0.6~GeV, where
an identification of the nonperturbative component with a set of
low-lying mesons appear natural, while sets 2 use a higher value,
2~GeV, where the validity of perturbation theory is better established.
The data are not good enough to allow a precise determination of
$\Lambda_{\mrm{QCD}}$. Therefore we use a fixed value
$\Lambda^{(4)} = 200$~MeV, in agreement with conventional
results for proton distributions. In the VMD component the $\rho^0$
and $\omega$ have been added coherently, so that
$\u\ubar : \d\dbar = 4 : 1$ at $Q_0$.
Unlike the $\p$, the $\gamma$ has a direct component where the photon
acts as an unresolved probe. In the definition of $F_2^{\gamma}$ this
adds a component $C^{\gamma}$, symbolically
\begin{equation}
F_2^{\gamma}(x,Q^2) = \sum_{\q} e_{\q}^2 \left[ f_{\q}^{\gamma} +
f_{\qbar}^{\gamma} \right] \otimes C_{\q} +
f_{\g}^{\gamma} \otimes C_{\g} + C^{\gamma} ~.
\end{equation}
Since $C^{\gamma} \equiv 0$ in leading order, and since we stay with
leading-order fits, it is permissible to neglect this complication.
Numerically, however, it makes a non-negligible difference. We
therefore make two kinds of fits, one DIS type with $C^{\gamma} = 0$
and one {\MSbar} type including the universal part
of $C^{\gamma}$.
When jet production is studied for real incoming photons, the standard
evolution approach is reasonable also for heavy flavours, i.e.\
predominantly the $\c$, but with a lower cut-off $Q_0 \approx m_{\c}$
for $\gamma \to \c\cbar$. Moving to Deeply Inelastic Scattering,
$\e\gamma \to \e X$, there is an extra kinematical constraint:
$W^2 = Q^2 (1-x)/x > 4 m_{\c}^2$. It is here better to use the
`Bethe-Heitler' cross section for $\gamma^* \gamma \to \c\cbar$.
Therefore each distribution appears in two variants. For applications
to real $\gamma$'s the parton distributions are calculated as the sum
of a vector-meson part and an anomalous part including all five flavours.
For applications to DIS, the sum runs over the same vector-meson part,
an anomalous part and possibly a $C^{\gamma}$ part for the three
light flavours, and a Bethe-Heitler part for $\c$ and $\b$.
In version 2 of the SaS distributions, which are the ones found here, the
extension from real to virtual photons was improved, and further options
made available \cite{Sch96}. The resolved components of the photon
are dampened by phenomenologically motivated virtuality-dependent
dipole factors, while the direct ones are explicitly calculable.
Thus eq.~(\ref{eq:decomprealga}) generalizes to
\begin{eqnarray}
f_a^{\gamma^\star}(x,Q^2,P^2)
& = & \sum_V \frac{4\pi\alphaem}{f_V^2} \left(
\frac{m_V^2}{m_V^2 + P^2} \right)^2 \,
f_a^{\gamma,V}(x,Q^2;\tilde{Q}_0^2)
\nonumber \\
& + & \frac{\alphaem}{2\pi} \, \sum_{\q} 2 e_{\q}^2 \,
\int_{Q_0^2}^{Q^2} \frac{{\d} k^2}{k^2} \, \left(
\frac{k^2}{k^2 + P^2} \right)^2 \, f_a^{\gamma,\q\qbar}(x,Q^2;k^2)
~,
\label{eq:decompvirtga}
\end{eqnarray}
with $P^2$ the photon virtuality and $Q^2$ the hard-process scale.
In addition to the introduction of the dipole form factors,
note that the lower input scale for the VMD states is here shifted
from $Q_0^2$ to some $\tilde{Q}_0^2 \geq Q_0^2$. This is based on
a study of the evolution equation \cite{Bor93} that shows that
the evolution effectively starts `later' in $Q^2$ for a virtual
photon. Equation~(\ref{eq:decompvirtga}) is one possible answer. It
depends on both $Q^2$ and $P^2$ in a non-trivial way, however, so that
results are only obtained by a time-consuming numerical integration
rather than as a simple parametrization. Therefore several other
alternatives are offered, that are in some sense equivalent, but can
be given in simpler form.
In addition to the SaS sets, {\Py} also contains the Drees--Grassie set
of parton distributions \cite{Dre85} and, as for the proton, there is
an interface to \tsc{Pdflib} and \tsc{LHAPDF}. These calls are
made with photon virtuality $P^2$ below the hard-process scale
$Q^2$. Further author-recommended constrains are implemented in the
interface to the GRS set \cite{Glu99} which, along with SaS, is among
the few also to define parton distributions of virtual photons.
However, these sets do not allow a subdivision of the photon parton
distributions into one VMD part and one anomalous part. This
subdivision is necessary a sophisticated modelling of $\gamma\p$ and
$\gamma\gamma$ events, see above and section \ref{sss:photoprod}.
As an alternative, for the VMD part alone, the
$\rho^0$ parton distribution can be found from the assumed equality
\begin{equation}
f^{\rho^0}_i = f^{\pi^0}_i = \frac{1}{2} \, (f^{\pi^+}_i +
f^{\pi^-}_i) ~.
\end{equation}
Thus any $\pi^+$ parton distribution set, from any library, can be
turned into a VMD $\rho^0$ set. The $\omega$ parton distribution is
assumed the same, while the $\phi$ and $\Jpsi$ ones are handled
in the very crude approximation
$f^{\phi}_{\s,\mrm{val}} = f^{\pi^+}_{\u,\mrm{val}}$ and
$f^{\phi}_{\mrm{sea}} = f^{\pi^+}_{\mrm{sea}}$.
The VMD part needs to be complemented by an
anomalous part to make up a full photon distribution. The latter
is fully perturbatively calculable, given the lower cut-off scale
$Q_0$. The SaS parameterization of the anomalous part is therefore used
throughout for this purpose. The $Q_0$ scale can be set freely in the
\ttt{PARP(15)} parameter.
The $f_i^{\gamma,\mrm{anom}}$ distribution can be further decomposed,
by the flavour and the $\pT$ of the original branching
$\gamma \to \q\qbar$. The flavour is distributed according to squared
charge (plus flavour thresholds for heavy flavours) and the $\pT$
according to $\d \pT^2 / \pT^2$ in the range $Q_0 < \pT < Q$.
At the branching scale, the photon only consists of a $\q\qbar$ pair,
with $x$ distribution $\propto x^2 + (1-x)^2$. A component
$f_a^{\gamma,\q\qbar}(x,Q^2;k^2)$, characterized by its $k \approx \pT$
and flavour, then is evolved homogeneously from $\pT$ to $Q$. For
theoretical studies it is convenient to be able to access a specific
component of this kind. Therefore also leading-order parameterizations
of these decomposed distributions are available \cite{Sch95}.
\subsubsection{Leptons}
\label{sss:estructfun}
Contrary to the hadron case, there is no necessity to introduce the
parton-distribution function concept for leptons. A lepton can be
considered as a point-like particle, with initial-state radiation
handled by higher-order matrix elements. However, the parton
distribution function approach offers a slightly simplified but very
economical description of initial-state radiation effects for any hard
process, also those for which higher-order corrections are not yet
calculated.
Parton distributions for electrons have been introduced in {\Py},
and are used also for muons and taus, with a trivial substitution
of masses. Alternatively, one is free to
use a simple `unresolved' $\e$, $f_{\e}^{\e}(x, Q^2) = \delta(x-1)$,
where the $\e$ retains the full original momentum.
Electron parton distributions are calculable entirely from first
principles, but different levels of approximation may be used.
The parton-distribution formulae in {\Py} are based on a
next-to-leading-order exponentiated description, see ref.
\cite{Kle89}, p. 34. The approximate behaviour is
\begin{eqnarray}
& & f_{\e}^{\e}(x,Q^2) \approx \frac{\beta}{2}
(1-x)^{\beta/2-1} ~; \nonumber \\
& & \beta = \frac{2 \alphaem}{\pi}
\left( \ln \frac{Q^2}{m_{\e}^2} -1 \right) ~.
\end{eqnarray}
The form is divergent but integrable for $x \to 1$, i.e.\ the electron
likes to keep most of the energy. To handle the numerical precision
problems for $x$ very close to unity, the parton distribution is set, by
hand, to zero for $x > 1-10^{-10}$, and is rescaled upwards in the range
$1-10^{-7} < x < 1-10^{-10}$, in such a way that the total area under the
parton distribution is preserved:
\begin{equation}
\left( f_{\e}^{\e}(x,Q^2) \right)_{\mrm{mod}} =
\left\{ \begin{array}{ll}
f_{\e}^{\e}(x,Q^2) & 0 \leq x \leq 1-10^{-7} \\[2mm]
\frac{\displaystyle 1000^{\beta/2}}{\displaystyle 1000^{\beta/2}-1}
\, f_{\e}^{\e}(x,Q^2) & 1-10^{-7} < x < 1-10^{-10} \\[4mm]
0 & x > 1-10^{-10} \, ~.
\end{array} \right.
\end{equation}
A separate issue is that electron beams may not be monochromatic,
more so than for other particles because of the small electron mass.
In storage rings the main mechanism is synchrotron radiation.
For future high-luminosity linear colliders, the beam--beam
interactions at the collision vertex may induce a quite significant
energy loss --- `beamstrahlung'. Note that neither of these are
associated with any off-shellness of the electrons, i.e.\ the resulting
spectrum only depends on $x$ and not $Q^2$. Examples of beamstrahlung
spectra are provided by the \tsc{Circe} program \cite{Ohl97}, with a
sample interface on the {\Py} webpages.
The branchings $\e \to \e \gamma$, which are responsible for the
softening of the $f_{\e}^{\e}$ parton distribution, also gives rise to
a flow of photons. In photon-induced hard processes, the
$f_{\gamma}^{\e}$ parton distribution can be used to describe the
equivalent flow of photons. In the next section, a complete
differential photon flux machinery is introduced. Here some simpler
first-order expressions are introduced, for the flux integrated up
to a hard interaction scale $Q^2$. There is some ambiguity in
the choice of $Q^2$ range over which emissions should be included.
The na\"{\i}ve (default) choice is
\begin{equation}
f_{\gamma}^{\e}(x,Q^2) = \frac{\alphaem}{2 \pi} \,
\frac{1+(1-x)^2}{x} \, \ln \left( \frac{Q^2}{m_{\e}^2} \right) ~.
\end{equation}
Here it is assumed that only one scale enters the problem, namely
that of the hard interaction, and that the scale of the branching
$\e \to \e \gamma$ is bounded from above by the hard interaction scale.
For a pure QCD or pure QED shower this is an appropriate procedure,
cf. section \ref{sss:showermatching}, but in other cases it may not be
optimal. In particular, for photoproduction the alternative that is
probably most appropriate is \cite{Ali88}:
\begin{equation}
f_{\gamma}^{\e}(x,Q^2) =
\frac{\alphaem}{2 \pi} \, \frac{1+(1-x)^2}{x}
\, \ln \left( \frac{Q^2_{\mmax} (1-x)}{m_{\e}^2 \, x^2} \right) ~.
\end{equation}
Here $Q^2_{\mmax}$ is a user-defined cut for the range of
scattered electron kinematics that is counted as
photoproduction. Note that we now deal with two different $Q^2$
scales, one related to the hard subprocess itself, which appears
as the argument of the parton distribution, and the other related to
the scattering of the electron, which is reflected in
$Q^2_{\mmax}$.
Also other sources of photons should be mentioned. One is the
beamstrahlung photons mentioned above, where again \tsc{Circe}
provides sample parameterizations. Another, particularly interesting
one, is laser backscattering, wherein an intense laser pulse is shot
at an incoming high-energy electron bunch. By Compton backscattering
this gives rise to a photon energy spectrum with a peak at a
significant fraction of the original electron energy \cite{Gin82}.
Both of these sources produce real photons, which can be considered as
photon beams of variable energy (see section \ref{ss:PYvaren}),
decoupled from the production process proper.
In resolved photoproduction or resolved $\gamma\gamma$
interactions, one has to include the parton distributions for quarks
and gluons inside the photon inside the electron. This is best done
with the machinery of the next section. However, as an older and
simpler alternative, $f_{\q,\g}^{\e}$ can be obtained by a numerical
convolution according to
\begin{equation}
f_{\q,\g}^{\e}(x, Q^2) =
\int_x^1 \frac{dx_{\gamma}}{x_{\gamma}} \,
f_{\gamma}^{\e}(x_{\gamma}, Q^2) \, f^{\gamma}_{\q,\g} \!
\left( \frac{x}{x_{\gamma}}, Q^2 \right) ~,
\label{pg:foldqgine}
\end{equation}
with $f^{\e}_{\gamma}$ as discussed above. The necessity for numerical
convolution makes this parton distribution evaluation rather slow
compared with the others; one should therefore only have it
switched on for resolved photoproduction studies.
One can obtain the positron distribution inside an electron,
which is also the electron sea parton distribution, by a convolution
of the two branchings $\e \to \e \gamma$ and $\gamma \to \ee$;
the result is \cite{Che75}
\begin{equation}
f_{\e^+}^{\e^-}(x,Q^2) =
\frac{1}{2} \, \left\{ \frac{\alphaem}{2 \pi} \,
\left( \ln \frac{Q^2}{m_{\e}^2} -1 \right) \right\}^2 \,
\frac{1}{x} \, \left( \frac{4}{3} - x^2 - \frac{4}{3} x^3 +
2x(1+x) \ln x \right) ~.
\end{equation}
Finally, the program also contains the distribution of a
transverse $\W^-$ inside an electron
\begin{equation}
f_{\W}^{\e}(x,Q^2) = \frac{\alphaem}{2 \pi} \,
\frac{1}{4 \ssintw} \, \frac{1+(1-x)^2}{x} \,
\ln \left( 1 + \frac{Q^2}{m_{\W}^2} \right) ~.
\end{equation}
\subsubsection{Equivalent photon flux in leptons}
\label{sss:equivgamma}
With the {\galep} option of a \ttt{PYINIT} call,
an $\e\p$ or $\e^+\e^-$ event (or corresponding processes with muons)
is factorized into the flux of virtual photons and the subsequent
interactions of such photons. While real photons always are transverse
(T), the virtual photons also allow a longitudinal (L) component.
This corresponds to cross sections
\begin{equation}
\d\sigma(\e\p\rightarrow\e\mbf{X}) = \sum_{\xi=\mathrm{T,L}}
\int \! \int \d y \, \d Q^2
\;f_{\gamma/\e}^{\xi}(y,Q^2)
\;\d\sigma(\gast_{\xi}\p\rightarrow\mbf{X})
\end{equation}
and
\begin{equation}
\d\sigma(\e\e\rightarrow\e\e\mbf{X}) = \!\!\!\sum_{\xi_1,\xi_2=\mathrm{T,L}}
\int \! \int \! \int \!\d y_1 \, \d Q_1^2 \, \d y_2 \, \d Q_2^2
\;f_{\gamma/\e}^{\xi_1}(y_1,Q_1^2) f_{\gamma/\e}^{\xi_2}(y_2,Q_2^2)
\;\d\sigma(\gast_{\xi_1}\gast_{\xi_2}\rightarrow\mbf{X})\;.
\end{equation}
For $\e\p$ events, this factorized ansatz is perfectly general, so long
as azimuthal distributions in the final state are not studied in detail.
In $\e^+\e^-$ events, it is not a good approximation when the
virtualities $Q_1^2$ and $Q_2^2$ of both photons become of the order of
the squared invariant mass $W^2$ of the colliding photons
\cite{Sch98}. In this region the cross section have terms that depend
on the relative azimuthal angle of the scattered leptons, and
the transverse and longitudinal polarizations are non-trivially mixed.
However, these terms are of order $Q_1^2Q_2^2/W^2$ and can
be neglected whenever at least one of the photons has low virtuality
compared to $W^2$.
When $Q^2/W^2$ is small, one can derive \cite{Bon73,Bud75,Sch98}
\begin{eqnarray}
f_{\gamma/l}^{\mathrm{T}}(y,Q^2) & = & \frac{\alphaem}{2\pi}
\left( \frac{(1+(1-y)^2}{y} \frac{1}{Q^2}-\frac{2m_{l}^2y}{Q^4}
\right)\;,\\
f_{\gamma/l}^{\mathrm{L}}(y,Q^2) & = & \frac{\alphaem}{2\pi}
\frac{2(1-y)}{y} \frac{1}{Q^2}\;,
\end{eqnarray}
where $l=\e^{\pm},~\mu^{\pm}$ or~$\tau^{\pm}$.
In $f_{\gamma/l}^{\mathrm{T}}$ the second term, proportional to
$m_{l}^2/Q^4$, is not leading log and is therefore often
omitted. Clearly it is irrelevant at large $Q^2$, but around the lower
cut-off $Q^2_{\mathrm{min}}$ it significantly dampens the small-$y$ rise
of the first term. (Note that $Q^2_{\mathrm{min}}$ is $y$-dependent,
so properly the dampening is in a region of the $(y,Q^2)$ plane.)
Overall, under realistic conditions, it reduces event rates by 5--10\%
\cite{Sch98,Fri93}.
The $y$ variable is defined as the light-cone fraction the photon takes
of the incoming lepton momentum. For instance, for $l^+l^-$ events,
\begin{equation}
y_i = \frac{q_i k_j}{k_i k_j} ~, \qquad j=2 (1)~\mathrm{for}~i=1 (2) ~,
\end{equation}
where the $k_i$ are the incoming lepton four-momenta and the $q_i$
the four-momenta of the virtual photons.
Alternatively, the energy fraction the photon takes in the rest frame
of the collision can be used,
\begin{equation}
x_i = \frac{q_i (k_1 + k_2)}{k_i (k_1 + k_2)} ~, \qquad i=1,2 ~.
\end{equation}
The two are simply related,
\begin{equation}
y_i = x_i + \frac{Q_i^2}{s} ~,
\end{equation}
with $s=(k_1 + k_2)^2$. (Here and in the following formulae we have
omitted the lepton and hadron mass terms when it is not of importance
for the argumentation.)
Since the Jacobian $\d(y_i, Q_i^2) / \d(x_i, Q_i^2) = 1$, either variable
would be an equally valid choice for covering the phase space.
Small $x_i$ values will be of less interest for us, since they lead
to small $W^2$, so $y_i/x_i \approx 1$ except in the high-$Q^2$ tail,
and often the two are used interchangeably. Unless special $Q^2$ cuts
are imposed, cross sections obtained with
$f_{\gamma/l}^{\mathrm{T,L}}(x,Q^2) \, \d x$ rather than
$f_{\gamma/l}^{\mathrm{T,L}}(y,Q^2) \, \d y$ differ only at the
per mil level. For comparisons with experimental cuts,
it is sometimes relevant to know which of the two is being used in
an analysis.
In the $\e\p$ kinematics, the $x$ and $y$ definitions give that
\begin{equation}
W^2 = x s = y s - Q^2 ~.
\end{equation}
The $W^2$ expression for $l^+l^-$ is more complicated, especially
because of the dependence on the relative azimuthal angle of the
scattered leptons,
$\varphi_{12} = \varphi_1 - \varphi_2$:
\begin{eqnarray}
W^2 & = & x_1 x_2 s + \frac{2 Q_1^2 Q_2^2}{s} -
2 \sqrt{1 - x_1 - \frac{Q_1^2}{s}} \sqrt{1 - x_2 - \frac{Q_2^2}{s}}
Q_1 Q_2 \cos\varphi_{12} \nonumber \\
& = & y_1 y_2 s - Q_1^2 - Q_2^2 + \frac{Q_1^2 Q_2^2}{s} -
2 \sqrt{1 - y_1} \sqrt{1 - y_2} Q_1 Q_2 \cos\varphi_{12} ~.
\label{W2gaga}
\end{eqnarray}
The lepton scattering angle $\theta_i$ is related to $Q_i^2$ as
\begin{equation}
Q_i^2 = \frac{x_i^2}{1-x_i} m_i^2 + (1-x_i) \left(
s - \frac{2}{(1-x_i)^2} m_i^2 - 2 m_j^2 \right) \sin^2(\theta_i/2) ~,
\end{equation}
with $m_i^2 = k_i^2 = {k'}^2_i$ and terms of $O(m^4)$ neglected.
The kinematical limits thus are
\begin{eqnarray}
(Q_i^2)_{\mathrm{min}} & \approx & \frac{x_i^2}{1 - x_i} m_i^2 ~, \\
(Q_i^2)_{\mathrm{max}} & \approx & (1 - x_i) s ~,
\end{eqnarray}
unless experimental conditions reduce the $\theta_i$ ranges.
In summary, we will allow the possibility of experimental cuts in the
$x_i$, $y_i$, $Q_i^2$, $\theta_i$ and $W^2$ variables. Within the
allowed region, the phase space is Monte Carlo sampled according to
$\prod_i (\d Q_i^2/Q_i^2) \, (\d x_i / x_i) \, \d \varphi_i$, with the
remaining flux factors combined with the cross section factors to give
the event weight used for eventual acceptance or rejection. This cross
section in its turn can contain the parton densities of a resolved
virtual photon, thus offering an effective convolution that gives
partons inside photons inside electrons.
\subsection{Kinematics and Cross Section for a Two-body Process}
\label{ss:kinemtwo}
In this section we begin the description of kinematics selection
and cross-section calculation. The example is for the case of a
$2 \to 2$ process, with final-state masses assumed to be vanishing.
Later on we will expand to finite fixed masses, and to resonances.
Consider two incoming beam particles in their c.m.\ frame, each with
energy $E_{\mrm{beam}}$. The total squared c.m.\ energy is then
$s = 4 E_{\mrm{beam}}^2$. The two partons that enter the hard
interaction do not carry the total beam momentum, but only fractions
$x_1$ and $x_2$, respectively, i.e.\ they have four-momenta
\begin{eqnarray}
p_1 & = & E_{\mrm{beam}}(x_1; 0, 0, x_1) ~, \nonumber \\
p_2 & = & E_{\mrm{beam}}(x_2; 0, 0, -x_2) ~.
\end{eqnarray}
There is no reason to put the incoming partons on the mass shell,
i.e.\ to have time-like incoming four-vectors,
since partons inside a particle are always virtual and thus
space-like. These space-like virtualities are introduced as
part of the initial-state parton-shower description, see section
\ref{sss:initshowtrans}, but do not affect the formalism of this
section, wherefore massless incoming partons is a sensible ansatz.
The one example where it would be appropriate to put a
parton on the mass shell is for an incoming lepton beam, but even
here the massless kinematics description is adequate as long as
the c.m.\ energy is correctly calculated with masses.
The squared invariant mass of the two partons is defined as
\begin{equation}
\hat{s} = (p_1 + p_2)^2 = x_1 \, x_2 \, s ~.
\end{equation}
Instead of $x_1$ and $x_2$, it is often customary to use $\tau$
and either $y$ or $x_{\mrm{F}}$:
\begin{eqnarray}
\tau & = & x_1 x_2 = \frac{\hat{s}}{s} ~; \\
y & = & \frac{1}{2} \ln \frac{x_1}{x_2} ~; \\
x_{\mrm{F}} & = & x_1 - x_2 ~.
\end{eqnarray}
In addition to $x_1$ and $x_2$, two additional variables are needed
to describe the kinematics of a scattering $1 + 2 \to 3 + 4$.
One corresponds to the azimuthal angle $\varphi$ of the scattering
plane around the beam axis. This angle is always isotropically
distributed for unpolarized incoming beam particles, and so need not
be considered further. The other variable can be picked as
$\hat{\theta}$, the polar angle of parton 3 in the c.m.\ frame of the
hard scattering. The conventional choice is to use the variable
\begin{equation}
\hat{t} = (p_1 - p_3)^2 = (p_2 - p_4)^2 =
- \frac{\hat{s}}{2} (1 - \cos \hat{\theta}) ~,
\end{equation}
with $\hat{\theta}$ defined as above. In the following, we will make
use of both $\hat{t}$ and $\hat{\theta}$. It is also customary
to define $\hat{u}$,
\begin{equation}
\hat{u} = (p_1 - p_4)^2 = (p_2 - p_3)^2 =
- \frac{\hat{s}}{2} (1 + \cos \hat{\theta}) ~,
\end{equation}
but $\hat{u}$ is not an independent variable since
\begin{equation}
\hat{s} + \hat{t} + \hat{u} = 0 ~.
\end{equation}
If the two outgoing particles have masses $m_3$ and $m_4$,
respectively, then the four-momenta in the c.m.\ frame of the hard
interaction are given by
\begin{equation}
\hat{p}_{3,4} = \left(
\frac{\hat{s} \pm (m_3^2-m_4^2)}{2\sqrt{\hat{s}}} ,
\pm \frac{\sqrt{\hat{s}}}{2} \, \beta_{34} \sin \hat{\theta}, 0,
\pm \frac{\sqrt{\hat{s}}}{2} \, \beta_{34} \cos \hat{\theta}
\right) ~,
\end{equation}
where
\begin{equation}
\beta_{34} = \sqrt{ \left( 1 - \frac{m_3^2}{\hat{s}} -
\frac{m_4^2}{\hat{s}} \right)^2 - 4 \, \frac{m_3^2}{\hat{s}} \,
\frac{m_4^2}{\hat{s}} } ~.
\label{pg:betafactor}
\end{equation}
Then $\hat{t}$ and $\hat{u}$ are modified to
\begin{equation}
\hat{t}, \hat{u} = - \frac{1}{2} \left\{ ( \hat{s} - m_3^2 - m_4^2 )
\mp \hat{s} \, \beta_{34} \cos \hat{\theta} \right\} ~,
\label{pg:thatuhat}
\end{equation}
with
\begin{equation}
\hat{s} + \hat{t} + \hat{u} = m_3^2 + m_4^2 ~.
\end{equation}
The cross section for the process $1 + 2 \to 3 + 4$ may be written as
\begin{eqnarray}
\sigma & = & \int \int \int \d x_1 \, \d x_2 \, \d \hat{t} \,
f_1(x_1, Q^2) \, f_2(x_2, Q^2) \, \frac{\d \hat{\sigma}}{\d \hat{t}}
\nonumber \\
& = & \int \int \int \frac{\d \tau}{\tau} \, \d y \, \d \hat{t} \,
x_1 f_1(x_1, Q^2) \, x_2 f_2(x_2, Q^2) \,
\frac{\d \hat{\sigma}}{\d \hat{t}} ~.
\label{pg:sigma}
\end{eqnarray}
The choice of $Q^2$ scale is ambiguous, and several alternatives are
available in the program. For massless outgoing particles the default
is the squared transverse momentum
\begin{equation}
Q^2 = \hat{p}_{\perp}^2 = \frac{\hat{s}}{4} \sin^2\hat{\theta} =
\frac{\hat{t}\hat{u}}{\hat{s}} ~,
\end{equation}
which is modified to
\begin{equation}
Q^2 = \frac{1}{2} (m_{\perp 3}^2 + m_{\perp 4}^2) =
\frac{1}{2} (m_3^2 + m_4^2) + \hat{p}_{\perp}^2 =
\frac{1}{2} (m_3^2 + m_4^2) +
\frac{\hat{t} \hat{u} - m_3^2 m_4^2}{\hat{s}}
\end{equation}
when masses are introduced in the final state. The mass term is
selected such that, for $m_3 = m_4 = m$, the expression reduces to
the squared transverse mass,
$Q^2 = \hat{m}_{\perp}^2 = m^2 + \hat{p}_{\perp}^2$.
For cases with space-like virtual incoming photons, of virtuality
$Q_i^2 = - m_i^2 = |p_i^2|$, a further generalization to
\begin{equation}
Q^2 = \frac{1}{2} (Q_1^2 + Q_2^2 + m_3^2 + m_4^2) + \hat{p}_{\perp}^2
\end{equation}
is offered.
The $\d \hat{\sigma}/\d \hat{t}$ expresses the differential
cross section for a scattering, as a function of the kinematical
quantities $\hat{s}$, $\hat{t}$ and $\hat{u}$, and of the relevant
masses. It is in this function that the physics of a given process
resides.
The performance of a machine is measured in terms of its
luminosity $\cal L$, which is directly proportional to the
number of particles in each bunch and to the bunch crossing
frequency, and inversely proportional to the area of the bunches at
the collision point. For a process with a $\sigma$ as given by
eq.~(\ref{pg:sigma}), the differential event rate is given
by $\sigma {\cal L}$, and the number of events collected
over a given period of time
\begin{equation}
N = \sigma \, \int {\cal L} \, \d t ~.
\end{equation}
The program does not calculate the number of events, but only the
integrated cross sections.
\subsection{Resonance Production}
\label{ss:kinemreson}
The simplest way to produce a resonance is by a $2 \to 1$ process.
If the decay of the resonance is not considered, the cross-section
formula does not depend on $\hat{t}$, but takes the form
\begin{equation}
\sigma = \int \int \frac{\d \tau}{\tau} \, \d y \,
x_1 f_1(x_1, Q^2) \, x_2 f_2(x_2, Q^2) \,
\hat{\sigma}(\hat{s}) ~.
\label{pg:sigmares}
\end{equation}
Here the physics is contained in the cross section
$\hat{\sigma}(\hat{s})$. The $Q^2$ scale is usually taken to
be $Q^2 = \hat{s}$.
In published formulae, cross sections are often given in
the zero-width approximation, i.e.\
$\hat{\sigma}(\hat{s}) \propto \delta (\hat{s} - m_R^2)$,
where $m_R$ is the mass of the resonance. Introducing the
scaled mass $\tau_R = m_R^2/s$, this corresponds to a delta
function $\delta (\tau - \tau_R)$, which can be used to
eliminate the integral over $\tau$.
However, what we normally want to do is replace the $\delta$
function by the appropriate Breit--Wigner shape. For a resonance
width $\Gamma_R$ this is achieved by the replacement
\begin{equation}
\delta (\tau - \tau_R) \to \frac{s}{\pi} \,
\frac{m_R \Gamma_R}{(s \tau - m_R^2)^2 + m_R^2 \Gamma_R^2} ~.
\label{pg:resshapeone}
\end{equation}
In this formula the resonance width $\Gamma_R$ is a constant.
An improved description of resonance shapes is obtained if
the width is made $\hat{s}$-dependent (occasionally also
referred to as mass-dependent width, since $\hat{s}$ is not
always the resonance mass), see e.g.\ \cite{Ber89}.
To first approximation, this means that the expression
$m_R \Gamma_R$ is to be replaced by $\hat{s} \Gamma_R / m_R$,
both in the numerator and the denominator. An intermediate
step is to perform this replacement only in the numerator.
This is convenient when not only $s$-channel resonance
production is simulated but also non-resonance $t$- or
$u$-channel graphs are involved, since mass-dependent widths
in the denominator here may give an imperfect cancellation of
divergences. (More about this below.)
To be more precise, in the program the quantity $H_R(\hat{s})$
is introduced, and the Breit--Wigner is written as
\begin{equation}
\delta (\tau - \tau_R) \to \frac{s}{\pi} \,
\frac{H_R(s \tau)}{(s \tau - m_R^2)^2 + H_R^2(s \tau)} ~.
\label{pg:resshapetwo}
\end{equation}
The $H_R$ factor is evaluated as a sum over all possible final-state
channels, $H_R = \sum_f H_R^{(f)}$. Each decay channel may have its
own $\hat{s}$ dependence, as follows.
A decay to a fermion pair, $R \to \f \fbar$, gives no contribution
below threshold, i.e.\ for $\hat{s} < 4 m_{\f}^2$. Above threshold,
$H_R^{(f)}$ is proportional to $\hat{s}$, multiplied by a threshold
factor $\beta (3 - \beta^2)/2$ for the vector part of a spin 1
resonance, by $\beta^3$ for the axial vector part, by $\beta^3$ for
a scalar resonance and by $\beta$ for a pseudoscalar one. Here
$\beta = \sqrt{1 - 4m_{\f}^2/\hat{s}}$.
For the decay into unequal masses, e.g.\ of the $\W^+$, corresponding
but more complicated expressions are used.
For decays into a quark pair, a first-order strong correction factor
$1 + \alphas(\hat{s}) / \pi$ is included in $H_R^{(f)}$. This is
the correct choice for all spin 1 colourless resonances, but is here
used for all resonances where no better knowledge is available.
Currently the major exception is top decay, where the
factor $1 - 2.5 \, \alphas(\hat{s}) / \pi$ is used to approximate
loop corrections \cite{Jez89}.
The second-order corrections are often known, but then are specific
to each resonance, and are not included. An option exists for the
$\gamma/\Z^0/\Z'^0$ resonances, where threshold effects due to
$\q\qbar$ bound-state formation are taken into account in a
smeared-out, average sense, see eq.~(\ref{pp:threshenh}).
For other decay channels, not into fermion pairs, the $\hat{s}$
dependence is typically more complicated. An example would be the
decay $\hrm^0 \to \W^+ \W^-$, with a nontrivial threshold and a subtle
energy dependence above that \cite{Sey95a}. Since a Higgs
with $m_{\hrm} < 2 m_{\W}$ could still decay in this channel, it is
in fact necessary to perform a two-dimensional integral over
the $W^{\pm}$ Breit--Wigner mass distributions to obtain the correct
result (and this has to be done numerically, at least in part).
Fortunately, a Higgs particle lighter than $2 m_{\W}$ is
sufficiently narrow that the integral only needs to be performed
once and for all at initialization (whereas most other partial
widths are recalculated whenever needed). Channels that proceed
via loops, such as $\hrm \to \g \g$, also display complicated
threshold behaviours.
The coupling structure within the electroweak sector is usually
(re)expressed in terms of gauge boson masses, $\alphaem$ and
$\ssintw$, i.e.\ factors of $G_{\F}$ are replaced according to
\begin{equation}
\sqrt{2} G_{\F} = \frac{\pi \, \alphaem}{\ssintw \, m_{\W}^2} ~.
\end{equation}
Having done that, $\alphaem$ is allowed to run \cite{Kle89},
and is evaluated at the $\hat{s}$ scale. Thereby the relevant
electroweak loop correction factors are recovered at the
$m_{\W}/m_{\Z}$ scale. However, the option exists to
go the other way and eliminate $\alphaem$ in favour of $G_{\F}$.
Currently $\ssintw$ is not allowed to run.
For Higgs particles and technipions, fermion masses enter not only
in the kinematics but also as couplings. The latter kind of quark masses
(but not the former, at least not in the program) are running with the
scale of the process, i.e.\ normally the resonance mass. The expression
used is \cite{Car96}
\begin{equation}
m(Q^2) = m_0 \left( \frac{\ln(k^2 m_0^2/\Lambda^2)}{\ln(Q^2/\Lambda^2)}
\right)^{12/(33 - 2 n_{\f})} ~.
\label{eq:runqmass}
\end{equation}
Here $m_0$ is the input mass at a reference scale $k m_0$, defined in the
{\MSbar} scheme. Typical choices are either $k=1$ or $k=2$; the latter
would be relevant if the reference scale is chosen at the $ \Q \Qbar$
threshold. Both $\Lambda$ and $n_{\f}$ are as given in $\alphas$.
In summary, we see that an $\hat{s}$ dependence may enter several
different ways into the $H_R^{(f)}$ expressions from which the
total $H_R$ is built up.
When only decays to a specific final state $f$ are considered, the
$H_R$ in the denominator remains the sum over all allowed decay
channels, but the numerator only contains the $H_R^{(f)}$ term
of the final state considered.
If the combined production and decay process $i \to R \to f$ is
considered, the same $\hat{s}$ dependence is implicit in the
coupling structure of $i \to R$ as one would have had in
$R \to i$, i.e.\ to first approximation there is a symmetry between
couplings of a resonance to the initial and to the final state.
The cross section $\hat{\sigma}$
is therefore, in the program, written in the form
\begin{equation}
\hat{\sigma}_{i \to R \to f}(\hat{s}) \propto \frac{\pi}{\hat{s}} \,
\frac{H_R^{(i)}(\hat{s}) \, H_R^{(f)}(\hat{s})}
{(\hat{s} - m_R^2)^2 + H_R^2(\hat{s})} ~.
\label{pg:Hinoutsym}
\end{equation}
As a simple example, the cross section for the process
$\e^- \br{\nu}_{\e} \to \W^- \to \mu^- \br{\nu}_{\mu}$
can be written as
\begin{equation}
\hat{\sigma}(\hat{s}) = 12 \, \frac{\pi}{\hat{s}} \,
\frac{H_{\W}^{(i)}(\hat{s}) \, H_{\W}^{(f)}(\hat{s})}
{(\hat{s} - m_{\W}^2)^2 + H_{\W}^2(\hat{s})} ~,
\end{equation}
where
\begin{equation}
H_{\W}^{(i)}(\hat{s}) = H_{\W}^{(f)}(\hat{s}) =
\frac{\alphaem(\hat{s})}{24 \, \ssintw} \, \hat{s} ~.
\end{equation}
If the effects of several initial and/or final states are studied,
it is straightforward to introduce an appropriate summation in the
numerator.
The analogy between the $H_R^{(f)}$ and $H_R^{(i)}$ cannot be pushed
too far, however. The two differ in several important aspects.
Firstly, colour factors appear reversed: the decay $R \to \q\qbar$
contains a colour factor $N_C = 3$ enhancement, while
$\q\qbar \to R$ is instead suppressed by a factor $1/N_C = 1/3$.
Secondly, the $1 + \alphas(\hat{s}) / \pi$ first-order correction
factor for the final state has to be replaced by a more complicated
$K$ factor for the initial state. This factor is not known usually, or
it is known (to first non-trivial order) but too lengthy to be included
in the program. Thirdly, incoming partons as a rule are space-like.
All the threshold suppression factors of the final-state expressions
are therefore irrelevant when production is considered. In sum, the
analogy between $H_R^{(f)}$ and $H_R^{(i)}$ is mainly useful as a
consistency cross-check, while the two usually are
calculated separately. Exceptions include the rather
messy loop structure involved in $\g\g \to \hrm^0$ and $\hrm^0 \to \g\g$,
which is only coded once.
It is of some interest to consider the observable resonance shape
when the effects of parton distributions are included. In a
hadron collider, to first approximation, parton distributions tend
to have a behaviour roughly like $f(x) \propto 1/x$ for small $x$
--- this is why $f(x)$ is replaced by $xf(x)$ in eq.~(\ref{pg:sigma}).
Instead, the basic parton-distribution behaviour
is shifted into the factor of $1/\tau$ in the integration phase space
$\d \tau/\tau$, cf. eq.~(\ref{pg:sigmares}). When convoluted with the
Breit--Wigner shape, two effects appear. One is that the overall
resonance is tilted: the low-mass tail is enhanced and the high-mass
one suppressed. The other is that an extremely long tail develops
on the low-mass side of the resonance: when $\tau \to 0$,
eq.~(\ref{pg:Hinoutsym}) with $H_R(\hat{s}) \propto \hat{s}$ gives
a $\hat{\sigma}(\hat{s}) \propto \hat{s} \propto \tau$, which
exactly cancels the $1/\tau$ factor mentioned above. Na\"{\i}vely, the
integral over $y$, $\int \d y = - \ln \tau$, therefore gives a
net logarithmic divergence of the resonance shape when $\tau \to 0$.
Clearly, it is then necessary to consider the shape of the
parton distributions in more detail. At not-too-small $Q^2$,
the evolution
equations in fact lead to parton distributions more strongly
peaked than $1/x$, typically with $xf(x) \propto x^{-0.3}$, and
therefore a divergence like $\tau^{-0.3}$ in the cross-section
expression. Eventually this divergence is regularized by a closing
of the phase space, i.e.\ that $H_R(\hat{s})$ vanishes faster than
$\hat{s}$, and by a less drastic small-$x$ parton-distribution
behaviour when $Q^2 \approx \hat{s} \to 0$.
The secondary peak at small $\tau$ may give a rather high
cross section, which can even rival that of the ordinary
peak around the nominal mass. This is the case, for instance,
with $\W$ production. Such a peak has never been observed
experimentally, but this is not surprising, since the background
from other processes is overwhelming at low $\hat{s}$.
Thus a lepton of one or a few GeV of transverse momentum is far
more likely to come from the decay of a charm or bottom hadron than
from an extremely off-shell $\W$ of a mass of a few GeV. When
resonance production is studied, it is therefore important to set
limits on the mass of the resonance, so as to agree with the
experimental definition, at least to first approximation. If not,
cross-section information given by the program may be very confusing.
Another problem is that often the matrix elements really are valid
only in the resonance region. The reason is that one usually includes
only the simplest $s$-channel graph in the calculation. It is this
`signal' graph that has a peak at the position of the resonance,
where it (usually) gives much larger cross sections than the other
`background' graphs. Away from the resonance position, `signal' and
`background' may be of comparable order, or the `background' may
even dominate. There is a quantum mechanical interference when some
of the `signal' and `background' graphs have the same initial and
final state, and this interference may be destructive or constructive.
When the interference is non-negligible, it is no longer meaningful
to speak of a `signal' cross section. As an example, consider the
scattering of longitudinal $\W$'s,
$\W^+_{\mrm{L}} \W^-_{\mrm{L}} \to \W^+_{\mrm{L}} \W^-_{\mrm{L}}$,
where the `signal' process is $s$-channel exchange of a Higgs.
This graph by itself is ill-behaved away from the resonance region.
Destructive interference with `background' graphs such as $t$-channel
exchange of a Higgs and $s$- and $t$-channel exchange of a $\gamma/\Z$
is required to save unitarity at large energies.
In $\ee$ colliders, the $f_{\e}^{\e}$ parton distribution is peaked
at $x = 1$ rather than at $x = 0$. The situation therefore is the
opposite, if one considers e.g.\ $\Z^0$ production in a machine
running at energies above $m_{\Z}$: the resonance-peak tail towards
lower masses is suppressed and the one towards higher masses enhanced,
with a sharp secondary peak at around the nominal energy of the
machine. Also in this case, an appropriate definition of cross
sections therefore is necessary --- with additional complications due
to the interference between $\gamma^*$ and $\Z^0$. When other processes
are considered, problems of interference with background appears also
here. Numerically the problems may be less pressing, however,
since the secondary peak is occurring in a high-mass region, rather
than in a more complicated low-mass one. Further, in $\ee$ there is
little uncertainty from the shape of the parton distributions.
In $2 \to 2$ processes where a pair of resonances are produced, e.g.
$\ee \to \Z^0 \hrm^0$, cross section are almost always given
in the zero-width approximation for the resonances. Here
two substitutions of the type
\begin{equation}
1 = \int \delta (m^2 - m_R^2) \, dm^2
\to \int \frac{1}{\pi} \,
\frac{m_R \Gamma_R}{(m^2 - m_R^2)^2 + m_R^2 \Gamma_R^2} \, dm^2
\label{pg:resshapethree}
\end{equation}
are used to introduce mass distributions
for the two resonance masses, i.e.\ $m_3^2$ and $m_4^2$.
In the formula, $m_R$ is the nominal mass and $m$ the actually
selected one. The
phase-space integral over $x_1$, $x_1$ and $\hat{t}$ in
eq.~(\ref{pg:sigma}) is then extended to involve also $m_3^2$ and
$m_4^2$. The effects of the mass-dependent width is only partly
taken into account, by replacing the nominal masses $m_3^2$ and
$m_4^2$ in the $\d \hat{\sigma} / \d \hat{t}$ expression by the
actually generated ones (also e.g.\ in the relation between
$\hat{t}$ and $\cos\hat{\theta}$), while the widths are evaluated
at the nominal masses. This is the equivalent of a simple replacement
of $m_R \Gamma_R$ by $\hat{s} \Gamma_R / m_R$ in the numerator of
eq.~(\ref{pg:resshapeone}), but not in the denominator. In addition, the
full threshold dependence of the widths, i.e.\ the velocity-dependent
factors, is not reproduced.
There is no particular reason why the full mass-dependence could not be
introduced, except for the extra work and time consumption needed for
each process. In fact, the matrix elements for several $\gammaZ$ and
$\W^{\pm}$ production processes do contain the full expressions.
On the other hand, the matrix elements given in the literature
are often valid only when the resonances are almost on the mass shell,
since some graphs have been omitted. As an example, the process
$\q\qbar \to \e^- \br{\nu}_{\e} \mu^+ \nu_{\mu}$ is dominated by
$\q\qbar \to \W^- \W^+$ when each of the two lepton pairs is close to
$m_{\W}$ in mass, but in general also receives contributions e.g.\ from
$\q\qbar \to \Z^0 \to \ee$, followed by $\e^+ \to \br{\nu}_{\e} \W^+$
and $\W^+ \to \mu^+ \nu_{\mu}$. The latter contributions are neglected
in cross sections given in the zero-width approximation.
Widths may induce gauge invariance problems, in particular when the
$s$-channel graph interferes with $t$- or $u$-channel ones. Then there
may be an imperfect cancellation of contributions at high energies,
leading to an incorrect cross section behaviour. The underlying reason
is that a Breit--Wigner corresponds to a resummation of terms of
different orders in coupling constants, and that therefore effectively
the $s$-channel contributions are calculated to higher orders than the
$t$- or $u$-channel ones, including interference contributions.
A specific example is $\e^+ \e^- \to \W^+ \W^-$, where $s$-channel
$\gamma^*/\Z^*$ exchange interferes with $t$-channel $\nu_{\e}$ exchange.
In such cases, a fixed width is used in the denominator. One could also
introduce procedures whereby the width is made to vanish completely at
high energies, and theoretically this is the cleanest, but the
fixed-width approach appears good enough in practice.
Another gauge invariance issue is when two particles of the same kind
are produced in a pair, e.g. $\g \g \to \t \tbar$. Matrix elements are
then often calculated for one common $m_{\t}$ mass, even though in real
life the masses $m_3 \neq m_4$. The proper gauge invariant procedure
to handle this would be to study the full six-fermion state obtained
after the two $\t \to \b \W \to \b \f_i \fbar_j$ decays, but that may
be overkill if indeed the $\t$'s are close to mass shell. Even when only
equal-mass matrix elements are available, Breit--Wigners are therefore
used to select two separate masses $m_3$ and $m_4$. From these two
masses, an average mass $\br{m}$ is constructed so that the
$\beta_{34}$ velocity factor of eq.~(\ref{pg:betafactor}) is retained,
\begin{equation}
\beta_{34}(\hat{s},\br{m}^2,\br{m}^2) =
\beta_{34}(\hat{s},m_3^2, m_4^2)
~~~\Rightarrow~~~
\br{m}^2 = \frac{m_3^2 + m_4^2}{2} -
\frac{(m_3^2 - m_4^2)^2}{4 \hat{s}}.
\end{equation}
This choice certainly is not unique, but normally should provide a sensible
behaviour, also around threshold.
Of course, the differential cross section is no longer guaranteed to
be gauge invariant when gauge bosons are involved, or positive definite.
The program automatically flags the latter situation as unphysical.
The approach may well break down when
either or both particles are far away from mass shell.
Furthermore, the preliminary choice of scattering
angle $\hat{\theta}$ is also retained. Instead of the correct $\hat{t}$
and $\hat{u}$ of eq.~(\ref{pg:thatuhat}), modified
\begin{equation}
\br{\hat{t}}, \br{\hat{u}} =
- \frac{1}{2} \left\{ ( \hat{s} - 2 \br{m}^2 )
\mp \hat{s} \, \beta_{34} \cos \hat{\theta} \right\} =
(\hat{t}, \hat{u}) - \frac{(m_3^2 - m_4^2)^2}{4 \hat{s}}
\end{equation}
can then be obtained. The $\br{m}^2$, $\br{\hat{t}}$ and
$\br{\hat{u}}$ are now used in the matrix elements to decide
whether to retain the event or not.
Processes with one final-state resonance and another ordinary
final-state product, e.g.\ $\q \g \to \W^+ \q'$, are treated in
the same spirit as the $2 \to 2$ processes with two resonances,
except that only one mass need be selected according to a
Breit--Wigner.
\subsection{Cross-section Calculations}
\label{ss:PYTcrosscalc}
In the program, the variables used in the generation of a
$2 \to 2$ process are $\tau$, $y$ and $z = \cos\hat{\theta}$.
For a $2 \to 1$ process, the $z$ variable can be integrated out,
and need therefore not be generated as part of the hard process,
except when the allowed angular range of decays is restricted.
In unresolved lepton beams, i.e.\ when
$f_{\e}^{\e}(x) = \delta(x-1)$, the variables $\tau$ and/or $y$
may be integrated out. We will cover all these special cases
towards the end of the section, and here concentrate on `standard'
$2 \to 2$ and $2 \to 1$ processes.
\subsubsection{The simple two-body processes}
In the spirit of section \ref{ss:MCdistsel}, we want to select simple
functions such that the true $\tau$, $y$ and $z$ dependence of the
cross sections is approximately modelled. In particular, (almost) all
conceivable kinematical peaks should be represented by separate
terms in the approximate formulae. If this can be achieved,
the ratio of the correct to the approximate cross sections will
not fluctuate too much, but allow reasonable Monte Carlo efficiency.
Therefore the variables are generated according to the distributions
$h_{\tau}(\tau)$, $h_y(y)$ and $h_z(z)$, where normally
\begin{eqnarray}
h_{\tau}(\tau) & = & \frac{c_1}{{\cal I}_1} \, \frac{1}{\tau} +
\frac{c_2}{{\cal I}_2} \, \frac{1}{\tau^2} +
\frac{c_3}{{\cal I}_3} \, \frac{1}{\tau(\tau + \tau_R)} +
\frac{c_4}{{\cal I}_4} \,
\frac{1}{(s\tau - m_R^2)^2 + m_R^2 \Gamma_R^2} \nonumber \\
& & + \frac{c_5}{{\cal I}_5} \, \frac{1}{\tau(\tau + \tau_{R'})} +
\frac{c_6}{{\cal I}_6} \,
\frac{1}{(s\tau - m_{R'}^2)^2 + m_{R'}^2 \Gamma_{R'}^2} ~, \\[1mm]
h_y(y) & = & \frac{c_1}{{\cal I}_1} \, (y - y_{\mmin}) +
\frac{c_2}{{\cal I}_2} \, (y_{\mmax} - y) +
\frac{c_3}{{\cal I}_3} \, \frac{1}{\cosh y} ~, \\[1mm]
h_z(z) & = & \frac{c_1}{{\cal I}_1} +
\frac{c_2}{{\cal I}_2} \, \frac{1}{a-z} +
\frac{c_3}{{\cal I}_3} \, \frac{1}{a+z} +
\frac{c_4}{{\cal I}_4} \, \frac{1}{(a-z)^2} +
\frac{c_5}{{\cal I}_5} \, \frac{1}{(a+z)^2} ~.
\label{pg:hforms}
\end{eqnarray}
Here each term is separately integrable, with an invertible primitive
function, such that generation of $\tau$, $y$ and $z$ separately
is a standard task, as described in section \ref{ss:MCdistsel}.
In the following we describe the details of the scheme, including
the meaning of the coefficients $c_i$ and ${\cal I}_i$, which are
separate for $\tau$, $y$ and $z$.
The first variable to be selected is $\tau$. The range of allowed
values, $\tau_{\mmin} \leq \tau \leq \tau_{\mmax}$, is generally
constrained by a number of user-defined requirements. A cut on the
allowed mass range is directly reflected in $\tau$, a cut on the
$\pT$ range indirectly. The first two terms
of $h_{\tau}$ are intended to represent a smooth $\tau$ dependence,
as generally obtained in processes which do not receive contributions
from $s$-channel resonances. Also $s$-channel exchange of
essentially massless particles ($\gamma$, $\g$, light quarks and
leptons) are accounted for, since these do not produce any separate
peaks at non-vanishing $\tau$. The last four terms of $h_{\tau}$ are
there to catch the peaks in the cross section from resonance
production. These terms are only included when needed. Each resonance
is represented by two pieces, a first to cover the interference with
graphs which peak at $\tau =0$, plus the variation of
parton distributions, and a second to approximate the
Breit--Wigner shape of the resonance itself. The subscripts $R$ and
$R'$ denote values pertaining to the two resonances, with
$\tau_R = m_R^2/s$. Currently there is only one process where the
full structure with two resonances is used, namely
$\f \fbar \to \gamma^*/\Z^0/\Z'^0$. Otherwise either one or no
resonance peak is taken into account.
The kinematically allowed range of $y$ values is
constrained by $\tau$, $|y| \leq - \frac{1}{2} \ln\tau$, and you
may impose additional cuts. Therefore the allowed range
$y_{\mmin} \leq y \leq y_{\mmax}$ is only constructed after
$\tau$ has been selected. The first two terms of $h_y$ give a fairly
flat $y$ dependence --- for processes which are symmetric in
$y \leftrightarrow -y$, they will add to give a completely flat $y$
spectrum between the allowed limits. In
principle, the natural subdivision would have been one term flat
in $y$ and one forward--backward asymmetric, i.e.\ proportional to
$y$. The latter is disallowed by the requirement of positivity,
however. The $y - y_{\mmin}$ and $y_{\mmax} - y$ terms
actually used give the same amount of freedom, but respect positivity.
The third term is peaked at around $y = 0$, and represents the bias
of parton distributions towards this region.
The allowed $z = \cos\hat{\theta}$ range is na\"{\i}vely
$-1 \leq z \leq 1$. However, most cross sections are divergent for
$z \to \pm 1$, so some kind of regularization is necessary. Normally
one requires $\pT \geq \pTmin$, which translates into
$z^2 \leq 1 - 4 \pTmin^2/(\tau s)$ for massless outgoing
particles. Since again the limits depend on $\tau$,
the selection of $z$ is done after that of
$\tau$. Additional requirements may constrain the range further.
In particular, a $p_{\perp\mmax}$ constraint may split the allowed
$z$ range into two, i.e.\ $z_{-\mmin} \leq z \leq z_{-\mmax}$ or
$z_{+\mmin} \leq z \leq z_{+\mmax}$. An un-split range is
represented by $z_{-\mmax} = z_{+\mmin} = 0$.
For massless outgoing particles
the parameter $a = 1$ in $h_z$, such that the five terms represent
a piece flat in angle and pieces peaked as $1/\hat{t}$, $1/\hat{u}$,
$1/\hat{t}^2$, and $1/\hat{u}^2$, respectively. For non-vanishing
masses one has $a = 1 + 2 m_3^2 m_4^2/\hat{s}^2$.
In this case, the full range $-1 \leq z \leq 1$ is therefore
available --- physically, the standard $\hat{t}$ and $\hat{u}$
singularities are regularized by the masses $m_3$ and $m_4$.
For each of the terms, the ${\cal I}_i$ coefficients represent the
integral over the quantity multiplying the coefficient $c_i$; thus,
for instance:
\begin{eqnarray}
h_{\tau}: & & {\cal I}_1 = \int \frac{\d \tau}{\tau} =
\ln \left( \frac{\tau_{\mmax}}{\tau_{\mmin}} \right) ~,
\nonumber \\
& & {\cal I}_2 = \int \frac{\d \tau}{\tau^2} =
\frac{1}{\tau_{\mmin}} - \frac{1}{\tau_{\mmax}} ~;
\nonumber \\
h_y: & & {\cal I}_1 = \int (y - y_{\mmin}) \, \d y =
\frac{1}{2} (y_{\mmax} - y_{\mmin})^2 ~; \nonumber \\
h_z: & & {\cal I}_1 = \int \d z = (z_{-\mmax} - z_{-\mmin}) +
(z_{+\mmax} - z_{+\mmin}) , \nonumber \\
& & {\cal I}_2 = \int \frac{\d z}{a-z} =
\ln \left( \frac{(a-z_{-\mmin})(a-z_{+\mmin})}
{(a-z_{-\mmax})(a-z_{-\mmin})} \right) ~.
\end{eqnarray}
The $c_i$ coefficients are normalized to unit sum for $h_{\tau}$,
$h_y$ and $h_z$ separately. They have a simple interpretation, as
the probability
for each of the terms to be used in the preliminary selection of
$\tau$, $y$ and $z$, respectively. The variation of the cross section
over the allowed phase space is explored in the initialization
procedure of a {\Py} run, and based on this knowledge the $c_i$
are optimized so as to give functions $h_{\tau}$, $h_y$ and $h_z$ that
closely follow the general behaviour of the true cross section.
For instance, the coefficient $c_4$ in $h_{\tau}$ is to be made larger
the more the total cross section is dominated by the region around
the resonance mass.
The phase-space points tested at initialization
are put on a grid, with the number of
points in each dimension given by the number of terms in the
respective $h$ expression, and with the position of each point
given by the median value of the distribution of one of the terms.
For instance, the $\d \tau / \tau$ distribution gives a median point at
$\sqrt{\tau_{\mmin}\tau_{\mmax}}$, and $\d \tau / \tau^2$ has the
median $2 \tau_{\mmin} \tau_{\mmax}
/ (\tau_{\mmin} + \tau_{\mmax})$.
Since the allowed $y$ and $z$ ranges depend on the $\tau$ value
selected, then so do the median points defined for these two
variables.
With only a limited set of phase-space points
studied at the initialization, the `optimal' set of coefficients
is not uniquely defined. To be on the safe side, 40\% of the total
weight is therefore assigned evenly between all allowed $c_i$,
whereas the remaining 60\% are assigned according to the relative
importance surmised, under the constraint that no coefficient is
allowed to receive a negative contribution from this second piece.
After a preliminary choice has been made of $\tau$, $y$ and $z$,
it is necessary to find the weight of the event, which is to be used
to determine whether to keep it or generate another one.
Using the relation $\d \hat{t} = \hat{s} \, \beta_{34} \, \d z / 2$,
eq.~(\ref{pg:sigma}) may be rewritten as
\begin{eqnarray}
\sigma & = & \int \int \int \frac{\d \tau}{\tau} \, \d y \,
\frac{\hat{s} \beta_{34}}{2} \d z \,
x_1 f_1(x_1, Q^2) \, x_2 f_2(x_2, Q^2) \,
\frac{\d \hat{\sigma}}{\d \hat{t}} \nonumber \\[1mm]
& = & \frac{\pi}{s} \int h_{\tau}(\tau) \, \d \tau
\int h_y(y) \, \d y \int h_z(z) \, \d z \, \beta_{34} \,
\frac{x_1 f_1(x_1, Q^2) \, x_2 f_2(x_2, Q^2)}
{\tau^2 h_{\tau}(\tau) \, h_y(y) \, 2 h_z(z)} \,
\frac{\hat{s}^2}{\pi} \frac{\d \hat{\sigma}}{\d \hat{t}}
\nonumber \\[1mm]
& = & \left\langle \frac{\pi}{s} \,
\frac{\beta_{34}}{\tau^2 h_{\tau}(\tau) \, h_y(y) \, 2 h_z(z)} \,
x_1 f_1(x_1, Q^2) \, x_2 f_2(x_2, Q^2) \,
\frac{\hat{s}^2}{\pi} \frac{\d \hat{\sigma}}{\d \hat{t}}
\right\rangle ~.
\label{pg:sigmamap}
\end{eqnarray}
In the middle line, a factor of $1 = h_{\tau}/h_{\tau}$
has been introduced to rewrite the $\tau$ integral
in terms of a phase space of unit volume:
$\int h_{\tau}(\tau) \, \d \tau = 1$ according to the relations
above. Correspondingly for the $y$ and $z$ integrals. In addition,
factors of $1 = \hat{s}/ (\tau s)$ and $1 = \pi / \pi$ are used to
isolate the dimensionless cross section
$(\hat{s}^2/\pi) \, \d \hat{\sigma} / \d \hat{t}$.
The content of the last line is that, with $\tau$, $y$ and $z$
selected according to the expressions $h_{\tau}(\tau)$, $h_y(y)$
and $h_z(z)$, respectively, the cross section is obtained as the
average of the final expression over all events. Since the $h$'s
have been picked to give unit volume, there is no need to multiply
by the total phase-space volume.
As can be seen, the cross section for a given Monte Carlo event is
given as the product of four factors, as follows:
\begin{Enumerate}
\item The $\pi/s$ factor, which is common to all events, gives the
overall dimensions of the cross section, in GeV$^{-2}$. Since
the final cross section is given in units of mb, the conversion
factor of 1 GeV$^{-2} = 0.3894$ mb is also included here.
\item Next comes the Jacobian, which compensates for the change
from the original to the final phase-space volume.
\item The parton-distribution function weight is obtained by making
use of the parton distribution libraries in {\Py} or externally.
The $x_1$ and $x_2$ values are obtained from $\tau$ and $y$ via the
relations $x_{1,2} = \sqrt{\tau} \exp(\pm y)$.
\item Finally, the dimensionless cross section
$(\hat{s}^2/\pi) \, \d \hat{\sigma} / \d \hat{t}$
is the quantity that has to be coded for each process separately,
and where the physics content is found.
\end{Enumerate}
Of course, the expression in the last line is not strictly necessary
to obtain the cross section by Monte Carlo integration. One could
also have used eq.~(\ref{pg:sigma}) directly, selecting phase-space
points evenly in $\tau$, $y$ and $\hat{t}$, and averaging over those
Monte Carlo weights. Clearly this would be much simpler, but the price
to be paid is that the weights of individual events could fluctuate
wildly. For instance, if the cross section contains a narrow
resonance, the few phase-space points that are generated in the
resonance region obtain large weights, while the rest do not.
With our procedure, a resonance would be included in the
$h_{\tau}(\tau)$ factor, so that more events would be generated
at around the appropriate $\tau_R$ value (owing to the $h_{\tau}$
numerator in the phase-space expression), but with these events
assigned a lower, more normal weight (owing to the factor
$1/h_{\tau}$ in the weight expression).
Since the weights fluctuate less, fewer phase-space points
need be selected to get a reasonable cross-section estimate.
In the program, the cross section is obtained as the average over all
phase-space points generated. The events actually handed on to
you should have unit weight, however (an option with weighted events
exists, but does not represent the mainstream usage). At
initialization, after the $c_i$ coefficients have been determined,
a search inside the allowed phase-space volume is therefore made
to find the maximum of the weight expression in the last line of
eq.~(\ref{pg:sigmamap}). In the subsequent generation of events,
a selected phase-space point is then retained with a probability
equal to the weight in the point divided by the maximum weight.
Only the retained phase-space points are considered further, and
generated as complete events.
The search for the maximum is begun by evaluating the weight in the
same grid of points as used to determine the $c_i$ coefficients.
The point with highest weight is used as starting point for a
search towards the maximum. In unfortunate cases, the convergence
could be towards a local maximum which is not the global one.
To somewhat reduce this risk, also the grid point with
second-highest weight is used for another search. After
initialization, when events are generated, a warning message
will be given by default at any time a phase-space
point is selected where the weight is larger than the maximum,
and thereafter the maximum weight is adjusted to reflect the new
knowledge. This means that events generated before this time have
a somewhat erroneous distribution in phase space, but if the
maximum violation is rather modest the effects should be negligible.
The estimation of the cross section is not affected by any of these
considerations, since the maximum weight does not enter into
eq.~(\ref{pg:sigmamap}).
For $2 \to 2$ processes with identical final-state particles,
the symmetrization factor of $1/2$ is explicitly included at the
end of the $\d \hat{\sigma} / \d \hat{t}$ calculation. In the final
cross section, a factor of 2 is retrieved because of integration
over the full phase space (rather than only half of it). That
way, no special provisions are needed in the phase-space
integration machinery.
\subsubsection{Resonance production}
We have now covered the simple $2 \to 2$ case. In a $2 \to 1$
process, the $\hat{t}$ integral is absent, and the differential
cross section $\d \hat{\sigma} / \d \hat{t}$ is replaced by
$\hat{\sigma}(\hat{s})$. The cross section may now be written as
\begin{eqnarray}
\sigma & = & \int \int \frac{\d \tau}{\tau} \, \d y \,
x_1 f_1(x_1, Q^2) \, x_2 f_2(x_2, Q^2) \,
\hat{\sigma}(\hat{s}) \nonumber \\
& = & \frac{\pi}{s} \int h_{\tau}(\tau) \, \d \tau
\int h_y(y) \, \d y \,
\frac{x_1 f_1(x_1, Q^2) \, x_2 f_2(x_2, Q^2)}
{\tau^2 h_{\tau}(\tau) \, h_y(y)} \,
\frac{\hat{s}}{\pi} \hat{\sigma}(\hat{s}) \nonumber \\
& = & \left\langle \frac{\pi}{s} \,
\frac{1}{\tau^2 h_{\tau}(\tau) \, h_y(y)} \,
x_1 f_1(x_1, Q^2) \, x_2 f_2(x_2, Q^2) \,
\frac{\hat{s}}{\pi} \hat{\sigma}(\hat{s})
\right\rangle ~.
\label{pg:sigmamapres}
\end{eqnarray}
The structure is thus exactly the same, but the $z$-related pieces
are absent, and the r\^ole of the dimensionless cross section is
played by $(\hat{s}/\pi) \hat{\sigma}(\hat{s})$.
If the range of allowed decay angles of the resonance is restricted,
e.g.\ by requiring the decay products to have a minimum transverse
momentum, effectively this translates into constraints on the
$z = \cos\hat{\theta}$ variable of the $2 \to 2$ process. The
difference is that the angular dependence of a resonance decay is
trivial, and that therefore the $z$-dependent factor can be easily
evaluated. For a spin-0 resonance, which decays isotropically, the
relevant weight is simply
$(z_{-\mmax} - z_{-\mmin})/2 + (z_{+\mmax} - z_{+\mmin})/2$.
For a transversely polarized spin-1 resonance the expression is,
instead,
\begin{equation}
\frac{3}{8}(z_{-\mmax} - z_{-\mmin}) +
\frac{3}{8}(z_{+\mmax} - z_{+\mmin}) +
\frac{1}{8}(z_{-\mmax} - z_{-\mmin})^3 +
\frac{1}{8}(z_{+\mmax} - z_{+\mmin})^3 ~.
\end{equation}
Since the allowed $z$ range could depend on $\tau$ and/or $y$ (it does
for a $\pT$ cut), the factor has to be evaluated for each individual
phase-space point and included in the expression of eq.
(\ref{pg:sigmamapres}).
For $2 \to 2$ processes where either of the final-state
particles is a resonance, or both, an additional choice has to
be made for
each resonance mass, eq.~(\ref{pg:resshapethree}). Since the allowed
$\tau$, $y$ and $z$ ranges depend on $m_3^2$ and $m_4^2$,
the selection of masses have to precede the choice of the other
phase-space variables. Just as for the other variables, masses
are not selected uniformly over the allowed range, but are rather
distributed according to a function $h_m(m^2) \, dm^2$, with a
compensating factor $1/h_m(m^2)$ in the Jacobian. The functional
form picked is normally
\begin{equation}
h_m(m^2) = \frac{c_1}{{\cal I}_1} \, \frac{1}{\pi} \,
\frac{m_R \Gamma_R}{(m^2 - m_R^2)^2 + m_R^2 \Gamma_R^2} +
\frac{c_2}{{\cal I}_2} +
\frac{c_3}{{\cal I}_3} \, \frac{1}{m^2} +
\frac{c_4}{{\cal I}_4} \, \frac{1}{m^4} ~.
\label{pg:reshdist}
\end{equation}
The definition of the ${\cal I}_i$ integrals is analogous to the one
before. The $c_i$ coefficients are not found by optimization, but
predetermined, normally to $c_1 = 0.8$, $c_2 = c_3 =0.1$, $c_4 = 0$.
Clearly, had the phase space and the cross section been independent
of $m_3^2$ and $m_4^2$, the optimal choice would have been to put
$c_1 = 1$ and have all other $c_i$ vanishing --- then the $1/h_m$
factor of the Jacobian would exactly have cancelled the
Breit--Wigner of eq.~(\ref{pg:resshapethree}) in the cross section.
The second and the third terms are there to cover the possibility
that the cross section does not die away quite as fast as given by
the na\"{\i}ve Breit--Wigner shape. In particular, the third term covers
the possibility of a secondary peak at small $m^2$, in a spirit
slightly similar to the one discussed for resonance production in
$2 \to 1$ processes.
The fourth term is only used for processes involving $\gammaZ$
production, where the $\gamma$ propagator
guarantees that the cross section does have a significant secondary
peak for $m^2 \to 0$. Therefore here the choice is $c_1 = 0.4$,
$c_2 = 0.05$, $c_3 = 0.3$ and $c_4 = 0.25$.
A few special tricks have been included to improve efficiency
when the allowed mass range of resonances is constrained by
kinematics or by user cuts. For instance, if a pair of equal
or charge-conjugate resonances are produced, such as in
$\ee \to \W^+ \W^-$, use is made of the constraint that the lighter
of the two has to have a mass smaller than half the c.m.\ energy.
\subsubsection{Lepton beams}
Lepton beams have to be handled slightly differently from what has
been described so far. One also has to distinguish between a lepton
for which parton distributions are included and one which is treated
as an unresolved point-like particle. The necessary modifications are
the same for $2 \to 2$ and $2 \to 1$ processes, however, since the
$\hat{t}$ degree of freedom is unaffected.
If one incoming beam is an unresolved lepton, the corresponding
parton-distribution piece collapses to a $\delta$ function. This
function can be used to integrate out the $y$ variable:
$\delta(x_{1,2} - 1) = \delta(y \pm (1/2) \ln \tau)$.
It is therefore only necessary to select the $\tau$ and the $z$
variables according to the proper distributions, with compensating
weight factors, and only one set of parton distributions has to be
evaluated explicitly.
If both incoming beams are unresolved leptons, both the $\tau$ and
the $y$ variables are trivially given: $\tau = 1$ and $y = 0$.
Parton-distribution weights disappear completely. For a $2 \to 2$
process, only the $z$ selection remains to be performed, while
a $2 \to 1$ process is completely specified, i.e.\ the cross section
is a simple number that only depends on the c.m.\ energy.
For a resolved electron, the $f_{\e}^{\e}$ parton distribution is
strongly peaked towards $x = 1$. This affects both the $\tau$
and the $y$ distributions, which are not well described by
either of the pieces in $h_{\tau}(\tau)$ or $h_y(y)$ in processes with
interacting $\e^{\pm}$. (Processes which involve e.g.\ the $\gamma$
content of the $\e$ are still well simulated, since
$f_{\gamma}^{\e}$ is peaked at small $x$.)
If both parton distributions are peaked close to 1, the
$h_{\tau}(\tau)$ expression in eq.~(\ref{pg:hforms}) is
therefore increased with one additional term of
the form $h_{\tau}(\tau) \propto 1 / (1 - \tau)$, with coefficients
$c_7$ and ${\cal I}_7$ determined as before. The divergence when
$\tau \to 1$ is cut off by our regularization procedure for the
$f_{\e}^{\e}$ parton distribution; therefore we only need consider
$\tau < 1 - 2 \times 10^{-10}$.
Correspondingly, the $h_y(y)$ expression is expanded with a term
$1/(1 - \exp(y-y_0))$ when incoming beam number 1 consists of a
resolved $\e^{\pm}$, and with a term $1/(1 - \exp(-y-y_0))$
when incoming beam number 2 consists of a resolved $\e^{\pm}$.
Both terms are present for an $\ee$ collider, only one for an
$\ep$ one. The coefficient $y_0 = - (1/2) \ln \tau$ is the na\"{\i}ve
kinematical limit of the $y$ range, $|y| < y_0$. From the
definitions of $y$ and $y_0$ it is easy to see
that the two terms above correspond to $1/(1-x_1)$ and $1/(1-x_2)$,
respectively, and thus are again regularized by our
parton-distribution function cut-off. Therefore the integration ranges
are $y < y_0 -10^{-10}$ for the first term and $y > - y_0 + 10^{-10}$
for the second one.
\subsubsection{Mixing processes}
\label{sss:mixingproc}
In the cross-section formulae given so far, we have deliberately
suppressed a summation over the allowed incoming flavours. For
instance, the process $\f\fbar \to \Z^0$ in a hadron collider
receives contributions from $\u\ubar \to \Z^0$, $\d\dbar \to \Z^0$,
$\s\sbar \to \Z^0$, and so on. These contributions share the
same basic form, but differ in the parton-distribution weights
and (usually) in a few coupling constants in the hard matrix
elements. It is therefore convenient to generate the terms
together, as follows:
\begin{Enumerate}
\item A phase-space point is picked, and all
common factors related to this choice are evaluated, i.e.\
the Jacobian and the common pieces of the matrix elements
(e.g.\ for a $\Z^0$ the basic Breit--Wigner shape, excluding
couplings to the initial flavour).
\item The parton-distribution-function library is called to produce
all the parton distributions, at the relevant $x$ and $Q^2$ values,
for the two incoming beams.
\item A loop is made over the two incoming flavours, one from each
beam particle. For each allowed set of incoming flavours, the full
matrix-element expression is constructed, using the common pieces and
the flavour-dependent couplings. This is multiplied by the common
factors and the parton-distribution weights to obtain a cross-section
weight.
\item Each allowed flavour combination is stored as a separate entry
in a table, together with its weight. In addition, a summed weight
is calculated.
\item The phase-space point is kept or rejected, according to a
comparison of the summed weight with the maximum weight obtained
at initialization. Also the cross-section Monte Carlo integration
is based on the summed weight.
\item If the point is retained, one of the allowed flavour
combinations is picked according to the relative weights stored
in the full table.
\end{Enumerate}
Generally, the flavours of the final state are either completely
specified by those of the initial state, e.g.\ as in $\q\g \to \q\g$,
or completely decoupled from them, e.g.\ as in
$\f\fbar \to \Z^0 \to \f'\fbar'$. In neither case need therefore the
final-state flavours be specified in the cross-section calculation.
It is only necessary, in the latter case, to include an overall
weight factor, which takes into account the summed contribution of
all final states that are to be simulated. For instance, if only
the process $\Z^0 \to \ee$ is studied, the relevant weight factor
is simply $\Gamma_{\e\e} / \Gamma_{\mrm{tot}}$. Once the kinematics
and the incoming flavours have been selected, the outgoing flavours
can be picked according to the appropriate relative probabilities.
In some processes, such as $\g\g \to \g\g$, several different colour
flows are allowed, each with its own kinematical dependence of the
matrix-element weight, see section \ref{sss:QCDjetclass}. Each colour
flow is then given as a separate entry in the table mentioned above,
i.e.\ in total an entry is characterized by the two incoming flavours,
a colour-flow index, and the weight. For an accepted phase-space
point, the colour flow is selected in the same way as the incoming
flavours.
The program can also allow the mixed generation of two or more
completely different processes, such as $\f\fbar \to \Z^0$ and
$\q\qbar \to \g\g$. In that case, each process is initialized
separately, with its own set of coefficients $c_i$ and so on.
The maxima obtained for the individual cross sections are all
expressed in the same units, even when the dimensionality of the
phase space is different. (This is because we always transform to
a phase space of unit volume,
$\int h_{\tau}(\tau) \, \d \tau \equiv 1$, etc.) The above
generation scheme need therefore only be generalized as follows:
\begin{Enumerate}
\item One process is selected among the allowed ones, with a relative
probability given by the maximum weight for this process.
\item A phase-space point is found, using the distributions
$h_{\tau}(\tau)$ and so on, optimized for this particular process.
\item The total weight for the phase-space point is evaluated,
again with Jacobians, matrix elements and allowed incoming flavour
combinations that are specific to the process.
\item The point is retained with a probability given by the ratio of
the actual to the maximum weight of the process. If the point is
rejected, one has to go back to step 1 and pick a new process.
\item Once a phase-space point has been accepted, flavours may be
selected, and the event generated in full.
\end{Enumerate}
It is clear why this works: although phase-space points are selected
among the allowed processes according to relative probabilities given
by the maximum weights, the probability that a point is accepted is
proportional to the ratio of actual to maximum weight. In total,
the probability for a given process to be retained is therefore only
proportional to the average of the actual weights, and any dependence
on the maximum weight is gone.
In $\gamma\p$ and $\gamma\gamma$ physics, the different components
of the photon give different final states, see section
\ref{sss:photoprod}. Technically, this introduces a further level
of administration, since each event class contains a set of (partly
overlapping) processes. From an ideological point of view, however,
it just represents one more choice to be made, that of event class,
before the selection of process in step 1 above. When a weighting
fails, both class and process have to be picked anew.
\subsection{Three- and Four-body Processes}
The {\Py} machinery to handle $2 \to 1$ and $2 \to 2$ processes
is fairly sophisticated and generic. The same cannot be said about
the generation of hard-scattering processes with more than two
final-state particles. The number of phase-space variables is
larger, and
it is therefore more difficult to find and transform away all possible
peaks in the cross section by a suitably biased choice of phase-space
points. In addition, matrix-element expressions for $2 \to 3$
processes are typically fairly lengthy. Therefore {\Py} only contains
a very limited number of $2 \to 3$ and $2 \to 4$ processes, and almost
each process is a special case of its own. It is therefore less
interesting to discuss details, and we only give a very generic
overview.
If the Higgs mass is not light, interactions among longitudinal
$\W$ and $\Z$ gauge bosons are of interest. In the program,
$2 \to 1$ processes such as $\W_{\mrm{L}}^+ \W_{\mrm{L}}^- \to \hrm^0$
and $2 \to 2$ ones such as
$\W_{\mrm{L}}^+ \W_{\mrm{L}}^- \to \Z_{\mrm{L}}^0 \Z_{\mrm{L}}^0$
are included.
The former are for use when the $\hrm^0$ still is reasonably narrow,
such that a resonance description is applicable, while the latter
are intended for high energies, where different contributions have
to be added up. Since the program does not contain
$\W_{\mrm{L}}$ or $\Z_{\mrm{L}}$
distributions inside hadrons, the basic hard scattering has
to be convoluted with the $\q \to \q' \W_{\mrm{L}}$ and
$\q \to \q \Z_{\mrm{L}}$
branchings, to yield effective $2 \to 3$ and $2 \to 4$ processes.
However, it is possible to integrate out the scattering angles of
the quarks analytically, as well as one energy-sharing variable
\cite{Cha85}. Only after an event has been accepted are these other
kinematical variables selected. This involves further choices of
random variables, according to a separate selection loop.
In total, it is therefore only necessary to introduce one additional
variable in the basic phase-space selection, which is chosen to be
$\hat{s}'$, the squared invariant mass of the full $2 \to 3$ or
$2 \to 4$ process, while $\hat{s}$ is used for the squared invariant
mass of the inner $2 \to 1$ or $2 \to 2$ process. The $y$ variable
is coupled to the full process, since parton-distribution weights
have to be given for the original quarks at
$x_{1,2} = \sqrt{\tau'} \exp{(\pm y)}$. The $\hat{t}$ variable is
related to the inner process, and thus not needed for the $2 \to 3$
processes. The selection of the $\tau' = \hat{s}'/s$ variable is
done after $\tau$, but before $y$ has been chosen. To improve the
efficiency, the selection is made according to a weighted phase space
of the form $\int h_{\tau'}(\tau') \, \d \tau'$, where
\begin{equation}
h_{\tau'}(\tau') = \frac{c_1}{{\cal I}_1} \frac{1}{\tau'} +
\frac{c_2}{{\cal I}_2} \, \frac{(1- \tau / \tau')^3}{\tau'^2} +
\frac{c_3}{{\cal I}_3} \, \frac{1}{\tau' (1 - \tau')} ~,
\end{equation}
in conventional notation. The $c_i$ coefficients are optimized
at initialization. The $c_3$ term, peaked at $\tau' \approx 1$,
is only used for $\ee$ collisions.
The choice of $h_{\tau'}$ is roughly matched to the
longitudinal gauge-boson flux factor, which is of the form
\begin{equation}
\left( 1 + \frac{\tau}{\tau'} \right) \,
\ln \left( \frac{\tau}{\tau'} \right) -
2 \left( 1 - \frac{\tau}{\tau'} \right) ~.
\end{equation}
For a light $\hrm$ the effective $\W$ approximation above breaks down,
and it is necessary to include the full structure of the
$\q \q' \to \q \q' \hrm^0$ (i.e.\ $\Z\Z$ fusion) and
$\q \q' \to \q'' \q''' \hrm^0$ (i.e.\ $\W\W$ fusion) matrix elements.
The $\tau'$, $\tau$ and $y$ variables are here retained, and selected
according to standard procedures. The Higgs mass is represented by the
$\tau$ choice; normally the $\hrm^0$ is so narrow that the $\tau$
distribution effectively collapses to a $\delta$ function. In addition,
the three-body final-state phase space is rewritten as
\begin{equation}
\left( \prod_{i=3}^5 \frac{1}{(2 \pi)^3} \frac{\d^3 p_i}{2 E_i}
\right) \, (2 \pi)^4 \delta^{(4)} (p_3 + p_4 + p_5 - p_1 - p_2) =
\frac{1}{(2 \pi)^5} \, \frac{\pi^2}{4 \sqrt{\lambda_{\perp 34}}}
\, \d p_{\perp 3}^2 \, \frac{\d \varphi_3}{2 \pi} \,
\d p_{\perp 4}^2 \, \frac{\d \varphi_4}{2 \pi} \, \d y_5 ~,
\end{equation}
where $\lambda_{\perp 34} = (m_{\perp 34}^2 - m_{\perp 3}^2 -
m_{\perp 4}^2)^2 - 4 m_{\perp 3}^2 m_{\perp 4}^2$.
The outgoing quarks are labelled 3 and 4, and the outgoing Higgs 5.
The $\varphi$ angles are selected isotropically, while the two
transverse momenta are picked, with some foreknowledge of the
shape of the $\W / \Z$ propagators in the cross sections, according
to $h_{\perp} (\pT^2) \, \d \pT^2$, where
\begin{equation}
h_{\perp}(\pT^2) = \frac{c_1}{{\cal I}_1} +
\frac{c_2}{{\cal I}_2} \, \frac{1}{m_R^2 + \pT^2} +
\frac{c_3}{{\cal I}_3} \, \frac{1}{(m_R^2 + \pT^2)^2} ~,
\end{equation}
with $m_R$ the $\W$ or $\Z$ mass, depending on process, and
$c_1 = c_2 = 0.05$, $c_3 = 0.9$. Within the limits given by the
other variable choices, the rapidity $y_5$ is chosen uniformly.
A final choice remains to be made, which comes from a twofold
ambiguity of exchanging the longitudinal momenta of partons 3
and 4 (with minor modifications if they are massive). Here
the relative weight can be obtained exactly from the form of the
matrix element itself.
\subsection{Resonance Decays}
\label{ss:resdecay}
Resonances (see section \ref{sss:resdecintro})
can be made to decay in two different routines. One is the
standard decay treatment (in \ttt{PYDECY}) that can be used
for any unstable particle, where decay channels
are chosen according to fixed probabilities, and decay angles
usually are picked isotropically in the rest frame of the resonance,
see section \ref{ss:partdecays}.
The more sophisticated treatment (in \ttt{PYRESD}) is the default
one for resonances produced in {\Py}, and is described here.
The ground rule is that everything in mass up to and including $\b$
hadrons is decayed with the simpler \ttt{PYDECY} routine, while
heavier particles are handled with \ttt{PYRESD}. This also includes
the $\gamma^* / \Z^0$, even though here the mass in principle could
be below the $\b$ threshold. Other resonances include, e.g., $\t$,
$\W^{\pm}$, $\hrm^0$, $\Z'^0$, $\W'^{\pm}$, $\H^0$, $\A^0$,
$\H^{\pm}$, and technicolor and supersymmetric particles.
\subsubsection{The decay scheme}
In the beginning of the decay treatment, either one or two
resonances may be present, the former represented by processes
such as $\q \qbar' \to \W^+$ and $\q \g \to \W^+ \q'$, the latter
by $\q \qbar \to \W^+ \W^-$. If the latter is the case, the
decay of the two resonances is considered in parallel
(unlike \ttt{PYDECY}, where one particle at a time is made to decay).
First the decay channel of each resonance is selected according to
the relative weights $H_R^{(f)}$, as described above, evaluated at
the actual mass of the resonance, rather than at the nominal one.
Threshold factors are therefore fully taken into account, with
channels automatically switched off below the threshold. Normally
the masses of the decay products are well-defined, but e.g.\ in
decays like $\hrm^0 \to \W^+ \W^-$ it is also necessary to select
the decay product masses. This is done according to two Breit--Wigners
of the type in eq.~(\ref{pg:resshapethree}), multiplied by the
threshold factor, which depends on both masses.
Next the decay angles of the resonance are selected isotropically in
its rest frame. Normally the full range of decay angles is available,
but in $2 \to 1$ processes the decay angles of the original resonance
may be restrained by user cuts, e.g.\ on the $\pT$ of the decay
products. Based on the angles, the four-momenta of the decay products
are constructed and boosted to the correct frame. As a rule, matrix
elements are given with quark and lepton masses assumed vanishing.
Therefore the four-momentum vectors constructed at this stage are
actually massless for all quarks and leptons.
The matrix elements may now be evaluated. For a process such as
$\q \qbar \to \W^+ \W^- \to \e^+ \nu_{\e} \mu^- \br{\nu}_{\mu}$,
the matrix element is a function of the four-momenta of the two
incoming fermions and of the four outgoing ones. An upper limit for
the event weight can be constructed from the cross section for the
basic process $\q \qbar \to \W^+ \W^-$, as already used to select
the two $\W$ momenta. If the weighting fails, new resonance decay
angles are picked and the procedure is iterated until acceptance.
Based on the accepted set of angles, the correct decay product
four-momenta are constructed, including previously neglected
fermion masses. Quarks and, optionally, leptons are allowed to
radiate, using the standard final-state showering machinery, with
maximum virtuality given by the resonance mass.
In some decays new resonances are produced, and these are then
subsequently allowed to decay. Normally only one resonance pair is
considered at a time, with the possibility of full correlations.
In a few cases triplets can also appear, but such configurations
currently are considered without inclusion of correlations. Also
note that, in a process like
$\q\qbar \to \Z^0 \hrm^0 \to \Z^0 \W^+ \W^- \to 6$ fermions,
the spinless nature of the $\hrm^0$ ensures that the $\W^{\pm}$
decays are decoupled from that of the $\Z^0$ (but not from each other).
\subsubsection{Cross-section considerations}
\label{sss:resdecaycross}
The cross section for a process which involves the production of one
or several resonances is always reduced to take into account channels
not allowed by user flags. This is trivial for a single $s$-channel
resonance, cf. eq.~(\ref{pg:Hinoutsym}), but can also be included
approximately if several layers of resonance decays are involved.
At initialization, the ratio between the user-allowed width and the
nominally possible one is evaluated and stored, starting from the
lightest resonances and moving upwards. As an example, one first finds
the reduction factors for $\W^+$ and for $\W^-$ decays, which need not
be the same if e.g.\ $\W^+$ is allowed to decay only to quarks and
$\W^-$ only to leptons. These factors enter together as a weight
for the $\hrm^0 \to \W^+ \W^-$ channel, which is thus reduced in
importance compared with other possible Higgs decay channels.
This is also reflected
in the weight factor of the $\hrm^0$ itself, where some channels are open
in full, others completely closed, and finally some (like the one
above) open but with reduced weight. Finally, the weight for the
process $\q\qbar \to \Z^0 \hrm^0$ is evaluated as the product of the
$\Z^0$ weight factor and the $\hrm^0$ one. The standard cross section
of the process is multiplied with this weight.
Since the restriction on allowed decay modes is already included in
the hard-process cross section, mixing of different event types is
greatly simplified, and the selection of decay channel chains is
straightforward. There is a price to be paid, however. The reduction
factors evaluated at initialization all refer to resonances at their
nominal masses. For instance, the $\W$ reduction factor is evaluated
at the nominal $\W$ mass, even when that factor is used, later on, in
the description of the decay of a 120 GeV Higgs, where at least one
$\W$ would be produced below this mass. We know of no case where this
approximation has any serious consequences, however.
The weighting procedure works because the number of resonances to be
produced, directly or in subsequent decays, can be derived
recursively already from the start. It does not work for particles
which could also be produced at later stages, such as the
parton-shower evolution and the fragmentation. For instance,
$\D^0$ mesons can be produced fairly late in the event generation
chain, in unknown numbers, and so weights could not be introduced
to compensate, e.g.\ for the forcing of decays only into $\pi^+ \K^-$.
One should note that this reduction factor is separate from the
description of the resonance shape itself, where the full
width of the resonance has to be used. This width is based on the
sum of all possible decay modes, not just the simulated ones.
{\Py} does allow the possibility to change also the underlying
physics scenario, e.g.\ to include the decay of a $\Z^0$ into a
fourth-generation neutrino.
Normally the evaluation of the reduction factors is straightforward.
However, for decays into a pair of equal or charge-conjugate
resonances, such as $\Z^0 \Z^0$ or $\W^+ \W^-$, it is possible to
pick combinations in such a way that the weight of the pair does
not factorize into a product of the weight of each resonance itself.
To be precise, any decay channel can be given seven different status
codes:
\begin{Itemize}
\item $-1$: a non-existent decay mode, completely switched off and of no
concern to us;
\item 0: an existing decay channel, which is switched off;
\item 1: a channel which is switched on;
\item 2: a channel switched on for particles, but off for antiparticles;
\item 3: a channel switched on for antiparticles, but off for particles;
\item 4: a channel switched on for one of the particles or antiparticles,
but not for both;
\item 5: a channel switched on for the other of the particles or
antiparticles, but not for both.
\end{Itemize}
The meaning of possibilities 4 and 5 is exemplified by the statement
`in a $\W^+\W^-$ pair, one $\W$ decays hadronically and the other
leptonically', which thus covers the cases where either $\W^+$ or
$\W^-$ decays hadronically.
Neglecting non-existing channels, each channel belongs to either of the
classes above. If we denote the total branching ratio into channels
of type $i$ by $r_i$, this then translates into the requirement
$r_0 + r_1 + r_2 + r_3 + r_4 + r_5 = 1$. For a single particle the
weight factor is $r_1 + r_2 + r_4$, and for a single antiparticle
$r_1 + r_3 + r_4$. For a pair of identical resonances, the joint weight
is instead
\begin{equation}
(r_1 + r_2)^2 + 2 (r_1 + r_2) (r_4 + r_5) + 2 r_4 r_5 ~,
\end{equation}
and for a resonance--antiresonance pair
\begin{equation}
(r_1 + r_2)(r_1 + r_3) + (2 r_1 + r_2 + r_3) (r_4 + r_5) +
2 r_4 r_5 ~.
\label{eq:WWallchancomb}
\end{equation}
If some channels come with a reduced weight because of restrictions
on subsequent decay chains, this may be described in terms of
properly reduced $r_i$, so that the sum is less than unity.
For instance, in a $\t\tbar \to \b\W^+ \, \bbar\W^-$ process,
the $\W$ decay modes may be restricted to $\W^+ \to \q\qbar$
and $\W^- \to \e^-\bar{\nu}_{\e}$, in which case
$(\sum r_i)_{\t} \approx 2/3$ and $(\sum r_i)_{\tbar} \approx 1/9$.
With index $\pm$ denoting resonance/antiresonance,
eq.~(\ref{eq:WWallchancomb}) then generalizes to
\begin{equation}
(r_1 + r_2)^+ (r_1 + r_3)^- + (r_1 + r_2)^+ (r_4 + r_5)^- +
(r_4 + r_5)^+ (r_1 + r_3)^- + r_4^+ r_5^- + r_5^+ r_4^- ~.
\label{eq:WWallchancombgen}
\end{equation}
\subsection{Nonperturbative Processes}
\label{ss:nonpertproc}
A few processes are not covered by the discussion so far. These are
the ones that depend on the details of hadronic wave functions,
and therefore are not strictly calculable perturbatively
(although perturbation theory may often provide some guidance).
What we have primarily in mind is elastic scattering, diffractive
scattering and low-$\pT$ `minimum-bias' events in hadron--hadron
collisions, but one can also find corresponding processes in
$\gamma \p$ and $\gamma \gamma$ interactions. The description
of these processes is rather differently structured from that of
the other ones, as is explained below. Models for
`minimum-bias' events are discussed in detail in sections
\ref{ss:multint}--\ref{ss:newmultint}, to which we refer for
details on this part of the program.
\subsubsection{Hadron--hadron interactions}
In hadron--hadron interactions, the total hadronic cross section
for $AB \to$ anything, $\sigma^{AB}_{\mrm{tot}}$, is calculated
using the parameterization of Donnachie and Landshoff \cite{Don92}.
In this approach, each cross section appears as the sum of one
pomeron term and one reggeon one
\begin{equation}
\sigma^{AB}_{\mrm{tot}}(s) = X^{AB} \, s^{\epsilon} +
Y^{AB} \, s^{-\eta} ~,
\label{pg:sigtotpomreg}
\end{equation}
where $s = E_{\mrm{cm}}^2$. The powers $\epsilon = 0.0808$ and
$\eta = 0.4525$ are expected to be universal, whereas the
coefficients $X^{AB}$ and $Y^{AB}$ are specific to each initial
state. (In fact, the high-energy behaviour given by the pomeron term
is expected to be the same for particle and antiparticle interactions,
i.e.\ $X^{\br{A}B} = X^{AB}$.) Parameterizations not provided
in \cite{Don92} have been calculated in the same spirit, making use
of quark counting rules \cite{Sch93a}.
The total cross section is subdivided according to
\begin{equation}
\sigma^{AB}_{\mrm{tot}}(s) = \sigma^{AB}_{\mrm{el}}(s) +
\sigma^{AB}_{\mrm{sd}(XB)}(s) + \sigma^{AB}_{\mrm{sd}(AX)}(s) +
\sigma^{AB}_{\mrm{dd}}(s) + \sigma^{AB}_{\mrm{nd}}(s) ~.
\label{pg:sigtotsplit}
\end{equation}
Here `el' is the elastic process $AB \to AB$, `sd$(XB)$' the single
diffractive $AB \to XB$, `sd$(AX)$' the single diffractive
$AB \to AX$, `dd' the double diffractive $AB \to X_1 X_2$,
and `nd' the non-diffractive ones. Higher diffractive topologies,
such as central diffraction, are currently neglected.
In the following, the elastic and diffractive cross sections
and event characteristics are described, as given in the model by
Schuler and Sj\"ostrand \cite{Sch94,Sch93a}. The non-diffractive
component is identified with the `minimum bias' physics already
mentioned, a practical but not unambiguous choice. Its cross section
is given by `whatever is left' according to eq.~(\ref{pg:sigtotsplit}),
and its properties are discussed in section \ref{ss:multint}.
At not too large squared momentum transfers $t$, the elastic cross
section can be approximated by a simple exponential fall-off. If one
neglects the small real part of the cross section, the optical
theorem then gives
\begin{equation}
\frac{\d\sigma_{\mrm{el}}}{\d t} =
\frac{\sigma_{\mrm{tot}}^2}{16 \pi} \, \exp(B_{\mrm{el}} t) ~,
\end{equation}
and $\sigma_{\mrm{el}} = \sigma_{\mrm{tot}}^2 / 16 \pi B_{\mrm{el}}$.
The elastic slope parameter is parameterized by
\begin{equation}
B_{\mrm{el}} = B^{AB}_{\mrm{el}}(s) = 2 b_A + 2 b_B +
4 s^{\epsilon} -4.2 ~,
\end{equation}
with $s$ given in units of GeV and $B_{\mrm{el}}$ in GeV$^{-2}$.
The constants $b_{A,B}$ are $b_{\p} = 2.3$,
$b_{\pi,\rho,\omega,\phi} = 1.4$, $b_{\J/\psi} = 0.23$.
The increase of the slope parameter with c.m.\ energy is faster
than the logarithmically one conventionally assumed; that way the
ratio $\sigma_{\mrm{el}} / \sigma_{\mrm{tot}}$ remains well-behaved
at large energies.
The diffractive cross sections are given by
\begin{eqnarray}
\frac{\d\sigma_{\mrm{sd}(XB)}(s)}{\d t \, \d M^2} & = &
\frac{g_{3\pomeron}}{16\pi} \, \beta_{A\pomeron} \,
\beta_{B\pomeron}^2 \, \frac{1}{M^2} \, \exp(B_{\mrm{sd}(XB)}t)
\, F_{\mrm{sd}} ~, \nonumber \\
\frac{\d\sigma_{\mrm{sd}(AX)}(s)}{\d t \, \d M^2} & = &
\frac{g_{3\pomeron}}{16\pi} \, \beta_{A\pomeron}^2 \,
\beta_{B\pomeron} \, \frac{1}{M^2} \, \exp(B_{\mrm{sd}(AX)}t)
\, F_{\mrm{sd}} ~, \nonumber \\
\frac{\d\sigma_{\mrm{dd}}(s)}{\d t \, \d M_1^2 \, \d M_2^2} & = &
\frac{g_{3\pomeron}^2}{16\pi} \, \beta_{A\pomeron} \,
\beta_{B\pomeron} \, \frac{1}{M_1^2} \, \frac{1}{M_2^2} \,
\exp(B_{\mrm{dd}}t) \, F_{\mrm{dd}} ~.
\end{eqnarray}
The couplings $\beta_{A\pomeron}$ are related to the pomeron term
$X^{AB} s^{\epsilon}$ of the total cross section parameterization,
eq.~(\ref{pg:sigtotpomreg}). Picking a reference scale
$\sqrt{s_{\mrm{ref}}} = 20$ GeV, the couplings are given by
$\beta_{A\pomeron}\beta_{B\pomeron} =
X^{AB} \, s_{\mrm{ref}}^{\epsilon}$. The triple-pomeron coupling is
determined from single-diffractive data to be
$g_{3\pomeron} \approx 0.318$ mb$^{1/2}$; within the context of the
formulae in this section.
The spectrum of diffractive masses $M$ is taken to begin
0.28 GeV $\approx 2 m_{\pi}$ above the mass of the respective
incoming particle and extend to the kinematical limit. The simple
$\d M^2 / M^2$ form is modified by the mass-dependence in the
diffractive slopes and in the $F_{\mrm{sd}}$ and $F_{\mrm{dd}}$
factors (see below).
The slope parameters are assumed to be
\begin{eqnarray}
B_{\mrm{sd}(XB)}(s) & = & 2b_B + 2\alpha' \ln\left(\frac{s}{M^2}\right)
~, \nonumber \\
B_{\mrm{sd}(AX)}(s) & = & 2b_A + 2\alpha' \ln\left(\frac{s}{M^2}\right)
~, \nonumber \\
B_{\mrm{dd}}(s) & = & 2\alpha' \ln\left(e^4 + \frac{s s_0}{M_1^2 M_2^2}
\right) ~.
\end{eqnarray}
Here $\alpha' = 0.25$ GeV$^{-2}$ and conventionally $s_0$ is picked as
$s_0 = 1 / \alpha'$. The term $e^4$ in $B_{\mrm{dd}}$ is added by hand
to avoid a breakdown of the standard expression for large values of
$M_1^2 M_2^2$. The $b_{A,B}$ terms protect $B_{\mrm{sd}}$ from breaking
down; however a minimum value of 2~GeV$^{-2}$ is still explicitly required
for $B_{\mrm{sd}}$, which comes into play e.g.\ for a $\Jpsi$ state
(as part of a VMD photon beam).
The kinematical range in $t$ depends on all the masses of the
problem. In terms of the scaled variables $\mu_1 = m_A^2/s$,
$\mu_2 = m_B^2/s$, $\mu_3 = M_{(1)}^2/s$ ($=m_A^2/s$ when $A$
scatters elastically), $\mu_4 = M_{(2)}^2/s$ ($=m_B^2/s$ when $B$
scatters elastically), and the combinations
\begin{eqnarray}
C_1 & = & 1 - (\mu_1 + \mu_2 + \mu_3 + \mu_4) +
(\mu_1 - \mu_2) (\mu_3 - \mu_4) ~, \nonumber \\
C_2 & = & \sqrt{(1 - \mu_1 -\mu_2)^2 - 4 \mu_1 \mu_2} \,
\sqrt{(1 - \mu_3 - \mu_4)^2 - 4 \mu_3 \mu_4} ~, \nonumber \\
C_3 & = & (\mu_3 - \mu_1) (\mu_4 - \mu_2) +
(\mu_1 + \mu_4 - \mu_2 - \mu_3) (\mu_1 \mu_4 - \mu_2 \mu_3) ~,
\end{eqnarray}
one has $t_{\mmin} < t < t_{\mmax}$ with
\begin{eqnarray}
t_{\mmin} & = & - \frac{s}{2} (C_1 + C_2) ~, \nonumber \\
t_{\mmax} & = & - \frac{s}{2} (C_1 - C_2)
= - \frac{s}{2} \, \frac{4C_3}{C_1 + C_2}
= \frac{s^2 C_3}{t_{\mmin}} ~.
\end{eqnarray}
The Regge formulae above for single- and double-diffractive events
are supposed to hold in certain asymptotic regions of the total phase
space. Of course, there will be diffraction also outside these
restrictive regions. Lacking a theory which predicts differential cross
sections at arbitrary $t$ and $M^2$ values, the Regge formulae are used
everywhere, but fudge factors are introduced in order to obtain
`sensible' behaviour in the full phase space. These factors are:
\begin{eqnarray}
F_{\mrm{sd}} & = & \left( 1 - \frac{M^2}{s} \right)
\left( 1 + \frac{c_{\mrm{res}} \, M_{\mrm{res}}^2}
{M_{\mrm{res}}^2 + M^2} \right) ~, \nonumber \\
F_{\mrm{dd}} & = &
\left( 1 - \frac{\left( M_1 + M_2 \right)^2}{s} \right)
\left( \frac{s\, m_{\p}^2}{ s\, m_{\p}^2 + M_1^2\, M_2^2} \right)
\nonumber \\
& \times &
\left( 1 + \frac{c_{\mrm{res}} \, M_{\mrm{res}}^2}
{M_{\mrm{res}}^2 + M_1^2} \right)
\left( 1 + \frac{c_{\mrm{res}} \, M_{\mrm{res}}^2}
{M_{\mrm{res}}^2 + M_2^2} \right) ~.
\end{eqnarray}
The first factor in either expression suppresses production close to
the kinematical limit. The second factor in $F_{dd}$ suppresses
configurations where the two diffractive systems overlap in rapidity
space. The final factors give an enhancement of the low-mass region,
where a resonance structure is observed in the data. Clearly a more
detailed modelling would have to be based on a set of exclusive states
rather than on this smeared-out averaging procedure. A reasonable fit
to $\p\p / \pbar\p$ data is obtained for
$c_{\mrm{res}} = 2$ and $M_{\mrm{res}} = 2$~GeV,
for an arbitrary particle $A$ which is diffractively excited
we use $M_{\mrm{res}}^A = m_A - m_{\p} + 2$~GeV.
The diffractive cross-section formulae above have been integrated
for a set of c.m.\ energies, starting at 10 GeV, and the results have
been parameterized. The form of these parameterizations is given in
ref. \cite{Sch94}, with explicit numbers for the $\p\p/\pbar\p$
case. {\Py} also contains similar parameterizations for
$\pi\p$ (assumed to be same as $\rho\p$ and $\omega\p$),
$\phi\p$, $\Jpsi\p$, $\rho\rho$ ($\pi\pi$ etc.), $\rho\phi$,
$\rho\Jpsi$, $\phi\phi$, $\phi\Jpsi$ and $\Jpsi\Jpsi$.
The processes above do not obey the ordinary event mixing strategy.
First of all, since their total cross sections are known, it is
possible to pick the appropriate process from the start, and then
remain with that choice. In other words, if the selection of
kinematical variables fails, one would not go back and pick a new
process, the way it was done in section \ref{sss:mixingproc}.
Second, it is not possible to impose any cuts or restrain allowed
%incoming or outgoing flavours: if not additional information were to
%be provided, it would make the whole scenario ill-defined.
%referee: confusing sentence
incoming or outgoing flavours; especially for minimum-bias events
the production at different transverse momenta is interrelated by
the underlying formalism.
Third, it is not recommended to mix generation of these processes
with that of any of the other ones: normally the other processes
have so small cross sections that they would almost never be
generated anyway. (We here exclude the cases of `underlying events'
and `pile-up events', where mixing is provided for, and even is a
central part of the formalism, see sections \ref{ss:multint} and
\ref{ss:pileup}.)
Once the cross-section parameterizations has been used to pick one
of the processes, the variables $t$ and $M$ are selected according
to the formulae given above.
A $\rho^0$ formed by $\gamma \to \rho^0$ in elastic or diffractive
scattering is polarized, and therefore its decay angular distribution
in $\rho^0 \to \pi^+ \pi^-$ is taken to be proportional to
$\sin^2 \theta$, where the reference axis is given by the $\rho^0$
direction of motion.
A light diffractive system, with a mass less than 1 GeV above the
mass of the incoming particle, is allowed to decay isotropically into
a two-body state. Single-resonance diffractive states, such as a
$\Delta^+$, are therefore not explicitly generated, but are assumed
described in an average, smeared-out sense.
A more massive diffractive system is subsequently treated
as a string with the quantum numbers of the original hadron. Since the
exact nature of the pomeron exchanged between the hadrons is unknown,
two alternatives are included. In the first, the pomeron is assumed to
couple to (valence) quarks, so that the string is stretched directly
between the struck quark and the remnant diquark (antiquark) of the
diffractive state. In the second, the interaction is rather with a
gluon, giving rise to a `hairpin' configuration in which the string
is stretched from a quark to a gluon and then back to a diquark
(antiquark). Both of these scenarios could be present in the data;
the default choice is to mix them in equal proportions.
There is experimental support for more complicated scenarios
\cite{Ing85}, wherein the pomeron has a partonic substructure,
which e.g.\ can lead to high-$\pT$ jet production in the diffractive
system. The full machinery, wherein a pomeron spectrum is convoluted
with a pomeron-proton hard interaction, is not available in {\Py}.
(But is found in the \tsc{PomPyt} program \cite{Bru96}.)
\subsubsection{Photoproduction and $\gamma\gamma$ physics}
\label{sss:photoprod}
The photon physics machinery in {\Py} has been largely expanded in
recent years. Historically, the model was first developed for
photoproduction, i.e.\ a real photon on a hadron target
\cite{Sch93,Sch93a}. Thereafter $\gamma\gamma$ physics was added
in the same spirit \cite{Sch94a,Sch97}. More recently also virtual
photons have been added to the description \cite{Fri00}, including
the nontrivial transition region between real photons and Deeply
Inelastic Scattering (DIS). In this section we partly trace
this evolution towards more complex configurations.
The total $\gamma\p$ and $\gamma\gamma$ cross sections can again be
parameterized in a
form like eq.~(\ref{pg:sigtotpomreg}), which is not so obvious since
the photon has more complicated structure than an ordinary hadron.
In fact, the structure is still not so well understood. The model we
outline is the one studied by Schuler and Sj\"ostrand
\cite{Sch93,Sch93a}, and further updated in \cite{Fri00}. In this model
the physical photon is represented by
\begin{equation}
| \gamma \rangle = \sqrt{Z_3} \, | \gamma_B \rangle +
\sum_{V=\rho^0,\omega,\phi,\Jpsi} \frac{e}{f_V} \, | V \rangle +
\sum_{\q} \frac{e}{f_{\q\qbar}} \, | \q\qbar \rangle +
\sum_{\ell=\e,\mu,\tau} \frac{e}{f_{\ell\ell}} \,
| \ell^+ \ell^- \rangle ~.
\label{pg:gammadecompo}
\end{equation}
By virtue of this superposition, one is led to a model of $\gamma\p$
interactions, where three different kinds of events may be
distinguished:
\begin{Itemize}
\item Direct events, wherein the bare photon $| \gamma_B \rangle$
interacts directly with a parton from the proton. The process is
perturbatively calculable, and no parton distributions of the photon
are involved. The typical event structure is two high-$\pT$ jets and
a proton remnant, while the photon does not leave behind any
remnant.
\item VMD events, in which the photon fluctuates into a vector meson,
predominantly a $\rho^0$. All the event classes known from ordinary
hadron--hadron interactions may thus occur here, such as elastic,
diffractive, low-$\pT$ and high-$\pT$ events. For the latter, one may
define (VMD) parton distributions of the photon, and the photon also
leaves behind a beam remnant. This remnant is smeared in transverse
momentum by a typical `primordial $k_{\perp}$' of a few hundred MeV.
\item Anomalous or GVMD (Generalized VMD) events, in which the photon
fluctuates into a $\q\qbar$ pair of larger virtuality than in the VMD
class. The initial parton distribution is perturbatively calculable,
as is the subsequent QCD evolution. It gives rise to the so-called
anomalous part of the parton distributions of the photon, whence one
name for the class. As long as only real photons were considered,
it made sense to define the cross section of this event class to be
completely perturbatively calculable, given some lower $\pT$ cut-off.
Thus only high-$\pT$ events could occur. However, alternatively, one
may view these states as excited higher resonances ($\rho'$ etc.),
thus the GVMD name. In this case one is lead to a picture which also
allows a low-$\pT$ cross section, uncalculable in perturbation theory.
The reality may well interpolate between these two extreme alternatives,
but the current framework more leans towards the latter point of view.
Either the $\q$ or the $\qbar$ plays the r\^ole of a beam remnant,
but this remnant has a larger $\pT$ than in the VMD case, related to
the virtuality of the $\gamma \leftrightarrow \q\qbar$ fluctuation.
\end{Itemize}
The $| \ell^+ \ell^- \rangle$ states can only interact strongly with
partons inside the hadron at higher orders, and can therefore be
neglected in the study of hadronic final states.
In order that the above classification is smooth and free of double
counting, one has to introduce scales that separate the three
components. The main one is $k_0$, which separates the low-mass
vector meson region from the high-mass $| \q\qbar \rangle$ one,
$k_0 \approx m_{\phi}/2 \approx 0.5$ GeV. Given this dividing
line to VMD states, the anomalous parton distributions are
perturbatively calculable. The total cross section of a state is not,
however, since this involves aspects of soft physics and eikonalization
of jet rates. Therefore an ansatz is chosen where the total cross section
of a state scales like $k_V^2/\kT^2$, where the adjustable parameter
$k_V \approx m_{\rho}/2$ for light quarks. The $\kT$ scale is
roughly equated with half the mass of the GVMD state.
The spectrum of GVMD states
is taken to extend over a range $k_0 < \kT < k_1$, where $k_1$ is
identified with the $\pTmin(s)$ cut-off of the perturbative jet
spectrum in hadronic interactions, $\pTmin(s) \approx 1.5$~GeV at
typical energies, see section \ref{ss:multint} and especially
eq.~(\ref{eq:ptmin}). Above that range, the states are assumed to be
sufficiently weakly interacting that no eikonalization procedure is
required, so that cross sections can be calculated perturbatively
without any recourse to pomeron phenomenology. There is some
arbitrariness in that choice, and some simplifications are required
in order to obtain a manageable description.
The VMD and GVMD/anomalous events are together called resolved ones.
In terms of high-$\pT$ jet production, the VMD and anomalous
contributions can be combined into a total resolved one, and the
same for parton-distribution functions. However, the two classes
differ in the structure of the underlying event and possibly in the
appearance of soft processes.
In terms of cross sections, eq.\ (\ref{pg:gammadecompo}) corresponds
to
\begin{equation}
\sigma_{\mrm{tot}}^{\gamma\p}(s) = \sigma_{\mrm{dir}}^{\gamma\p}(s) +
\sigma_{\mrm{VMD}}^{\gamma\p}(s) + \sigma_{\mrm{anom}}^{\gamma\p}(s) ~.
\label{pg:gammacrossdecompo}
\end{equation}
The direct cross section is, to lowest order, the perturbative cross
section for the two processes $\gamma\q \to \q\g$ and
$\gamma\g \to \q\qbar$, with a lower cut-off $\pT > k_1$, in order to
avoid double-counting with the interactions of the GVMD states.
Properly speaking, this should be multiplied by the $Z_3$ coefficient,
\begin{equation}
Z_3 = 1 -
\sum_{V=\rho^0,\omega,\phi,\Jpsi} \left( \frac{e}{f_V} \right)^2 -
\sum_{\q} \left( \frac{e}{f_{\q\qbar}} \right)^2 -
\sum_{\ell=\e,\mu,\tau} \left( \frac{e}{f_{\ell\ell}} \right)^2 ~,
\end{equation}
but normally $Z_3$ is so close to unity as to make no difference.
The VMD factor $(e/f_V)^2 = 4\pi\alphaem/f_V^2$ gives the probability
for the transition $\gamma \to V$. The coefficients $f_V^2/4\pi$ are
determined from data to be (with a non-negligible amount of
uncertainty) 2.20 for $\rho^0$, 23.6 for $\omega$, 18.4 for $\phi$
and 11.5 for $\Jpsi$. Together these numbers imply that the photon
can be found in a VMD state about 0.4\% of the time, dominated by the
$\rho^0$ contribution. All the properties of the VMD interactions
can be obtained by appropriately scaling down $V\p$ physics
predictions. Thus the whole machinery developed in the previous
section for hadron--hadron interactions is directly applicable.
Also parton distributions of the VMD component inside the photon
are obtained by suitable rescaling.
The contribution from the `anomalous' high-mass fluctuations to the
total cross section is obtained by a convolution of the fluctuation
rate
\begin{equation}
\sum_{\q} \left( \frac{e}{f_{\q\qbar}} \right)^2 \approx
\frac{\alphaem}{2\pi} \, \left( 2 \sum_{\q} e_{\q}^2 \right)
\int_{k_0}^{k_1} \frac{\d \kT^2}{\kT^2} ~,
\label{anomintfirst}
\end{equation}
which is to be multiplied by the abovementioned reduction factor
$k_V^2/\kT^2$ for the total cross section, and all scaled by the
assumed real vector meson cross section.
As an illustration of this scenario, the phase space of $\gamma\p$
events may be represented by a $(\kT,\pT)$ plane.
Two transverse momentum scales are distinguished: the
photon resolution scale $\kT$ and the hard interaction scale $\pT$.
Here $\kT$ is a measure of the virtuality of a fluctuation of the
photon and $\pT$ corresponds to the most virtual rung of the ladder,
possibly apart from $\kT$.
As we have discussed above, the low-$\kT$ region corresponds to
VMD and GVMD states that encompasses both perturbative high-$\pT$ and
nonperturbative low-$\pT$ interactions. Above $k_1$, the region is split
along the line $\kT = \pT$. When $\pT > \kT$ the photon is resolved by
the hard interaction, as described by the anomalous part of the photon
distribution function. This is as in the GVMD sector, except that we should
(probably) not worry about multiple parton--parton interactions. In the
complementary region $\kT > \pT$, the $\pT$ scale is just part of the
traditional evolution of the parton distributions of the proton up to
the scale of $\kT$, and
thus there is no need to introduce an internal structure of the photon.
One could imagine the direct class of events as extending below $k_1$
and there being the low-$\pT$ part of the GVMD class, only appearing
when a hard interaction at a larger $\pT$ scale would not preempt it.
This possibility is implicit in the standard cross section framework.
In $\gamma\gamma$ physics \cite{Sch94a,Sch97}, the superposition in
eq.~(\ref{pg:gammadecompo}) applies separately for each of the two
incoming photons. In total there are therefore $3 \times 3 = 9$
combinations. However, trivial symmetry reduces this to six distinct
classes, written in terms of the total cross section
(cf. eq.~(\ref{pg:gammacrossdecompo})) as
\begin{eqnarray}
\sigma_{\mrm{tot}}^{\gamma\gamma}(s) & = &
\sigma_{\mrm{dir}\times\mrm{dir}}^{\gamma\gamma}(s) +
\sigma_{\mrm{VMD}\times\mrm{VMD}}^{\gamma\gamma}(s) +
\sigma_{\mrm{GVMD}\times\mrm{GVMD}}^{\gamma\gamma}(s) \nonumber \\
& + & 2 \sigma_{\mrm{dir}\times\mrm{VMD}}^{\gamma\gamma}(s) +
2 \sigma_{\mrm{dir}\times\mrm{GVMD}}^{\gamma\gamma}(s) +
2 \sigma_{\mrm{VMD}\times\mrm{GVMD}}^{\gamma\gamma}(s) ~.
\label{pg:gagacrossdecompo}
\end{eqnarray}
A parameterization of the total $\gamma\gamma$ cross section is found in
\cite{Sch94a,Sch97}.
The six different kinds of $\gamma\gamma$ events are thus:
\begin{Itemize}
\item The direct$\times$direct events, which correspond to the
subprocess $\gamma\gamma \to \q\qbar$ (or $\ell^+\ell^-$).
The typical event structure is two high-$\pT$ jets and no beam
remnants.
\item The VMD$\times$VMD events, which have the same properties as
the VMD $\gamma\p$ events. There are four by four combinations of
the two incoming vector mesons, with one VMD factor for each meson.
\item The GVMD$\times$GVMD events, wherein each photon
fluctuates into a $\q\qbar$ pair of larger virtuality than in the
VMD class. The `anomalous' classification assumes that
one parton of each pair gives a beam remnant, whereas
the other (or a daughter parton thereof) participates in a high-$\pT$
scattering. The GVMD concept implies the presence also of low-$\pT$
events, like for VMD.
\item The direct$\times$VMD events, which have the same properties as
the direct $\gamma\p$ events.
\item The direct$\times$GVMD events, in which a bare photon
interacts with a parton from the anomalous photon.
The typical structure is then two high-$\pT$ jets and a beam remnant.
\item The VMD$\times$GVMD events, which have the same properties
as the GVMD $\gamma\p$ events.
\end{Itemize}
Like for photoproduction events, this can be illustrated in a
parameter space, but now three-dimensional, with axes
given by the $k_{\perp 1}$, $k_{\perp 2}$ and $p_{\perp}$ scales. Here
each $k_{\perp i}$ is a measure of the virtuality of a fluctuation of
a photon, and $p_{\perp}$ corresponds to the most virtual rung on
the ladder between the two photons, possibly excepting the endpoint
$k_{\perp i}$ ones. So, to first approximation, the coordinates along the
$k_{\perp i}$ axes determine the characters of the interacting photons
while $p_{\perp}$ determines the character of the interaction process.
Double-counting should be avoided by trying to impose a consistent
classification. Thus, for instance, $p_{\perp} > k_{\perp i}$
with $k_{\perp 1} < k_0$ and $k_0 < k_{\perp 2} < k_1$ gives a hard
interaction between a VMD and a GVMD photon, while
$k_{\perp 1} > p_{\perp} > k_{\perp 2}$ with $k_{\perp 1} > k_1$
and $k_{\perp 2} < k_0$ is a single-resolved process
(direct$\times$VMD; with $p_{\perp}$ now in the parton distribution
evolution).
In much of the literature, where a coarser classification is used,
our direct$\times$direct is called direct, our direct$\times$VMD
and direct$\times$GVMD is called single-resolved since they both
involve one resolved photon which gives a beam remnant,
and the rest are called double-resolved since both photons are resolved
and give beam remnants.
If the photon is virtual, it has a reduced probability to fluctuate into
a vector meson state, and this state has a reduced interaction probability.
This can be modelled by a traditional dipole factor
$(m_V^2/(m_V^2 + Q^2))^2$ for a photon of virtuality $Q^2$, where
$m_V \to 2 \kT$ for a GVMD state. Putting it all together, the cross
section of the GVMD sector of photoproduction then scales like
\begin{equation}
\int_{k_0^2}^{k_1^2} \frac{\d\kT^2}{\kT^2} \, \frac{k_V^2}{\kT^2} \,
\left( \frac{4\kT^2}{4\kT^2 + Q^2} \right)^2 ~.
\end{equation}
For a virtual photon the DIS process $\gast \q \to \q$ is also possible,
but by gauge invariance its cross section must vanish in the limit
$Q^2 \to 0$. At large $Q^2$, the direct processes can be considered
as the $\mathcal{O}(\alphas)$ correction to the lowest-order
DIS process, but the direct ones survive for $Q^2 \to 0$. There is no
unique prescription for a proper combination at all $Q^2$, but we have
attempted an approach that gives the proper limits and minimizes
double-counting. For large $Q^2$, the DIS $\gast\p$ cross section
is proportional to the structure function $F_2 (x, Q^2)$ with the
Bjorken $x = Q^2/(Q^2 + W^2)$. Since normal parton distribution
parameterizations are frozen below some $Q_0$ scale and therefore do not
obey the gauge invariance condition, an ad hoc factor
$(Q^2/(Q^2 + m_{\rho}^2))^2$ is introduced for the conversion from
the parameterized $F_2(x,Q^2)$ to a $\sigma_{\mrm{DIS}}^{\gast\p}$:
\begin{equation}
\sigma_{\mrm{DIS}}^{\gast\p} \simeq
\left( \frac{Q^2}{Q^2 + m_{\rho}^2} \right)^2 \,
\frac{4\pi^2\alphaem}{Q^2} F_2(x,Q^2) =
\frac{4\pi^2\alphaem Q^2}{(Q^2+m_{\rho}^2)^2} \,
\sum_{\q} e_{\q}^2 \, \left\{ x q(x, Q^2) + x \br{q}(x,Q^2) \right\}
~.
\label{eq:sigDIS}
\end{equation}
Here $m_{\rho}$ is some nonperturbative hadronic mass parameter,
for simplicity identified with the $\rho$ mass. One of the
$Q^2/(Q^2+m_{\rho}^2)$ factors is required already to give finite
$\sigma_{\mrm{tot}}^{\gamma\p}$ for conventional parton distributions,
and could be viewed as a screening of the individual partons at small
$Q^2$. The second factor is chosen to give not only a finite but
actually a vanishing $\sigma_{\mrm{DIS}}^{\gast\p}$
for $Q^2 \to 0$ in order to retain the pure photoproduction description
there. This latter factor thus is more a matter of convenience, and
other approaches could have been pursued.
In order to avoid double-counting between DIS and direct events, a
requirement $\pT > \max(k_1, Q)$ is imposed on direct events. In the
remaining DIS ones, denoted lowest order (LO) DIS, thus $\pT < Q$.
This would suggest a subdivision
$\sigma_{\mrm{LO\,DIS}}^{\gast\p} = \sigma_{\mrm{DIS}}^{\gast\p} -
\sigma_{\mrm{direct}}^{\gast\p}$, with $\sigma_{\mrm{DIS}}^{\gast\p}$
given by eq.~(\ref{eq:sigDIS}) and $\sigma_{\mrm{direct}}^{\gast\p}$
by the perturbative matrix elements. In the limit $Q^2 \to 0$, the
DIS cross section is now constructed to vanish while the direct is not,
so this would give $\sigma_{\mrm{LO\,DIS}}^{\gast\p} < 0$. However,
here we expect the correct answer not to be a negative number but an
exponentially suppressed one, by a Sudakov form factor. This modifies
the cross section:
\begin{equation}
\sigma_{\mrm{LO\,DIS}}^{\gast\p} = \sigma_{\mrm{DIS}}^{\gast\p} -
\sigma_{\mrm{direct}}^{\gast\p}
~~ \longrightarrow ~~
\sigma_{\mrm{DIS}}^{\gast\p} \; \exp \left( - \frac{%
\sigma_{\mrm{direct}}^{\gast\p}}{\sigma_{\mrm{DIS}}^{\gast\p}} \right) \;.
\label{eq:LODIS}
\end{equation}
Since we here are in a region where the DIS cross section is no longer the
dominant one, this change of the total DIS cross section is not essential.
The overall picture, from a DIS perspective, now requires three scales
to be kept track of. The traditional DIS region is the strongly ordered
one, $Q^2 \gg \kT^2 \gg \pT^2$, where DGLAP-style evolution
\cite{Alt77, Gri72} is responsible for the event
structure. As always, ideology wants strong ordering, while
the actual classification is based on ordinary ordering
$Q^2 > \kT^2 > \pT^2$. The region $\kT^2 > \max(Q^2,\pT^2)$ is also
DIS, but of the $\mathcal{O}(\alphas)$ direct kind. The region
where $\kT$ is the smallest scale corresponds to
non-ordered emissions, that then go beyond DGLAP validity,
while the region $\pT^2 > \kT^2 > Q^2$ cover the interactions of a
resolved virtual photon. Comparing with the plane of real
photoproduction, we conclude that the whole region
$\pT > \kT$ involves no double-counting, since we have made no
attempt at a non-DGLAP DIS description but can choose to cover this
region entirely by the VMD/GVMD descriptions. Actually, it is only
in the corner $\pT < \kT < \min(k_1, Q)$ that an overlap can occur
between the resolved
and the DIS descriptions. Some further considerations show that
usually either of the two is strongly suppressed in this region,
except in the range of intermediate $Q^2$ and rather small $W^2$.
Typically, this is the region where $x \approx Q^2/(Q^2 + W^2)$ is not
close to zero, and where $F_2$ is dominated by the valence-quark
contribution. The latter behaves roughly $\propto (1-x)^n$, with an
$n$ of the order of 3 or 4. Therefore we will introduce a corresponding
damping factor to the VMD/GVMD terms.
In total, we have now arrived at our ansatz for all $Q^2$:
\begin{equation}
\sigma_{\mrm{tot}}^{\gast\p} =
\sigma_{\mrm{DIS}}^{\gast\p} \; \exp \left( -
\frac{\sigma_{\mrm{direct}}^{\gast\p}}{\sigma_{\mrm{DIS}}^{\gast\p}}
\right) + \sigma_{\mrm{direct}}^{\gast\p} +
\left( \frac{W^2}{Q^2 + W^2} \right)^n \left(
\sigma_{\mrm{VMD}}^{\gast\p} +
\sigma_{\mrm{GVMD}}^{\gast\p} \right) \;,
\label{eq:gammapallQ}
\end{equation}
with four main components. Most of these in their turn have
a complicated internal structure, as we have seen.
Turning to $\gast\gast$ processes, finally, the parameter space is now
five-dimensional: $Q_1$, $Q_2$, $k_{\perp 1}$, $k_{\perp 2}$ and $\pT$.
As before, an effort is made to avoid double-counting, by having a
unique classification of each region in the five-dimensional space.
Remaining double-counting is dealt with as above.
In total, our ansatz for $\gast\gast$ interactions at all $Q^2$ contains
13 components: 9 when two VMD, GVMD or direct photons interact, as is
already allowed for real photons, plus a further 4 where a `DIS photon'
from either side interacts with a VMD or GVMD one. With the label
resolved used to denote VMD and GVMD, one can write
\begin{eqnarray}
\sigma_{\mrm{tot}}^{\gast\gast} (W^2, Q_1^2, Q_2^2) & = &
\sigma_{\mrm{DIS}\times\mrm{res}}^{\gast\gast} \;
\exp \left( - \frac{\sigma_{\mrm{dir}\times\mrm{res}}^{\gast\gast}}%
{\sigma_{\mrm{DIS}\times\mrm{res}}^{\gast\gast}} \right) +
\sigma_{\mrm{dir}\times\mrm{res}}^{\gast\gast}
\nonumber \\
& + & \sigma_{\mrm{res}\times\mrm{DIS}}^{\gast\gast} \;
\exp \left( - \frac{\sigma_{\mrm{res}\times\mrm{dir}}^{\gast\gast}}%
{\sigma_{\mrm{res}\times\mrm{DIS}}^{\gast\gast}} \right) +
\sigma_{\mrm{res}\times\mrm{dir}}^{\gast\gast}
\label{eq:gagaallQ} \\
& + & \sigma_{\mrm{dir}\times\mrm{dir}}^{\gast\gast}
+ \left( \frac{W^2}{Q_1^2 + Q_2^2 + W^2} \right)^3 \;
\sigma_{\mrm{res}\times\mrm{res}}^{\gast\gast} \nonumber
\end{eqnarray}
Most of the 13 components in their turn have a complicated internal
structure, as we have seen.
An important note is that the $Q^2$ dependence of the DIS and direct
photon interactions is implemented in the matrix element expressions,
i.e.\ in processes such as $\gast\gast \to \q\qbar$ or
$\gast \q \to \q \g$ the photon virtuality explicitly enters. This is
different from VMD/GVMD, where dipole factors are used to reduce the
total cross sections and the assumed flux of
partons inside a virtual photon relative to those of a real one, but
the matrix elements themselves contain no dependence on the virtuality
either of the partons or of the photon itself.
Typically results are obtained with the SaS~1D parton distributions
for the virtual transverse photons
\cite{Sch95,Sch96}, since these are well matched to our framework, e.g.
allowing a separation of the VMD and GVMD/anomalous components.
Parton distributions of virtual longitudinal photons are by default
given by some $Q^2$-dependent factor times the transverse ones.
The set by Ch\'yla \cite{Chy00} allows more precise modelling
here, but indications are that many studies will not be sensitive
to the detailed shape.
The photon physics machinery is of considerable complexity, and
so the above is only a brief summary. Further details can be found
in the literature quoted above. Some topics are also covered in
other places in this manual, e.g. the flux of transverse and
longitudinal photons in section \ref{sss:equivgamma}, scale choices
for parton density evaluation in section \ref{ss:kinemtwo}, and
further aspects of the generation machinery and switches in section
\ref{ss:photoanddisclass}.
\clearpage
\begin{table}[pt]
\caption{Subprocess codes, part 1. First column is `+' for processes
implemented and blank for those that are/were only foreseen. Second is
the subprocess number {\ISUB}, and third the description of the
process. The final column gives references from which the
cross sections have been obtained. See text for further information.
\protect\label{t:procone} }
\begin{center}
\begin{tabular}{|c|r|l|l|}
\hline
In & No. & Subprocess & Reference \\
\hline
+ & 1 & $\f_i \fbar_i \to \gammaZ$ & \cite{Eic84} \\
+ & 2 & $\f_i \fbar_j \to \W^+$ & \cite{Eic84} \\
+ & 3 & $\f_i \fbar_i \to \hrm^0$ & \cite{Eic84} \\
& 4 & $\gamma \W^+ \to \W^+$ & \\
+ & 5 & $\Z^0 \Z^0 \to \hrm^0$ & \cite{Eic84,Cha85} \\
& 6 & $\Z^0 \W^+ \to \W^+$ & \\
& 7 & $\W^+ \W^- \to \Z^0$ & \\
+ & 8 & $\W^+ \W^- \to \hrm^0$ & \cite{Eic84,Cha85} \\
+ & 10 & $\f_i \f_j \to \f_k \f_l$ (QFD) & \cite{Ing87a} \\
+ & 11 & $\f_i \f_j \to \f_i \f_j$ (QCD) &
\cite{Com77,Ben84,Eic84} \\
+ & 12 & $\f_i \fbar_i \to \f_k \fbar_k$ &
\cite{Com77,Ben84,Eic84} \\
+ & 13 & $\f_i \fbar_i \to \g \g$ & \cite{Com77,Ben84} \\
+ & 14 & $\f_i \fbar_i \to \g \gamma$ & \cite{Hal78,Ben84} \\
+ & 15 & $\f_i \fbar_i \to \g \Z^0$ & \cite{Eic84} \\
+ & 16 & $\f_i \fbar_j \to \g \W^+$ & \cite{Eic84} \\
& 17 & $\f_i \fbar_i \to \g \hrm^0$ & \\
+ & 18 & $\f_i \fbar_i \to \gamma \gamma$ & \cite{Ber84} \\
+ & 19 & $\f_i \fbar_i \to \gamma \Z^0$ & \cite{Eic84} \\
+ & 20 & $\f_i \fbar_j \to \gamma \W^+$ & \cite{Eic84,Sam91} \\
& 21 & $\f_i \fbar_i \to \gamma \hrm^0$ & \\
+ & 22 & $\f_i \fbar_i \to \Z^0 \Z^0$ & \cite{Eic84,Gun86} \\
+ & 23 & $\f_i \fbar_j \to \Z^0 \W^+$ & \cite{Eic84,Gun86} \\
+ & 24 & $\f_i \fbar_i \to \Z^0 \hrm^0$ & \cite{Ber85} \\
+ & 25 & $\f_i \fbar_i \to \W^+ \W^-$ & \cite{Bar94,Gun86} \\
+ & 26 & $\f_i \fbar_j \to \W^+ \hrm^0$ & \cite{Eic84} \\
& 27 & $\f_i \fbar_i \to \hrm^0 \hrm^0$ & \\
+ & 28 & $\f_i \g \to \f_i \g$ & \cite{Com77,Ben84} \\
+ & 29 & $\f_i \g \to \f_i \gamma$ & \cite{Hal78,Ben84} \\
+ & 30 & $\f_i \g \to \f_i \Z^0$ & \cite{Eic84} \\
+ & 31 & $\f_i \g \to \f_k \W^+$ & \cite{Eic84} \\
+ & 32 & $\f_i \g \to \f_i \hrm^0$ & \cite{Bar88} \\
+ & 33 & $\f_i \gamma \to \f_i \g$ & \cite{Duk82} \\
+ & 34 & $\f_i \gamma \to \f_i \gamma$ & \cite{Duk82} \\
+ & 35 & $\f_i \gamma \to \f_i \Z^0$ & \cite{Gab86} \\
+ & 36 & $\f_i \gamma \to \f_k \W^+$ & \cite{Gab86} \\
& 37 & $\f_i \gamma \to \f_i \hrm^0$ & \\
\hline
\end{tabular}
\end{center}
\end{table}
\begin{table}[pt]
\caption{Subprocess codes, part 2. Comments as before.
\protect\label{t:proctwo} }
\begin{center}
\begin{tabular}{|c|r|l|l|}
\hline
In & No. & Subprocess & Reference \\
\hline
& 38 & $\f_i \Z^0 \to \f_i \g$ & \\
& 39 & $\f_i \Z^0 \to \f_i \gamma$ & \\
& 40 & $\f_i \Z^0 \to \f_i \Z^0$ & \\
& 41 & $\f_i \Z^0 \to \f_k \W^+$ & \\
& 42 & $\f_i \Z^0 \to \f_i \hrm^0$ & \\
& 43 & $\f_i \W^+ \to \f_k \g$ & \\
& 44 & $\f_i \W^+ \to \f_k \gamma$ & \\
& 45 & $\f_i \W^+ \to \f_k \Z^0$ & \\
& 46 & $\f_i \W^+ \to \f_k \W^+$ & \\
& 47 & $\f_i \W^+ \to \f_k \hrm^0$ & \\
& 48 & $\f_i \hrm^0 \to \f_i \g$ & \\
& 49 & $\f_i \hrm^0 \to \f_i \gamma$ & \\
& 50 & $\f_i \hrm^0 \to \f_i \Z^0$ & \\
& 51 & $\f_i \hrm^0 \to \f_k \W^+$ & \\
& 52 & $\f_i \hrm^0 \to \f_i \hrm^0$ & \\
+ & 53 & $\g \g \to \f_k \fbar_k$ & \cite{Com77,Ben84} \\
+ & 54 & $\g \gamma \to \f_k \fbar_k$ & \cite{Duk82} \\
& 55 & $\g \Z^0 \to \f_k \fbar_k$ & \\
& 56 & $\g \W^+ \to \f_k \fbar_l$ & \\
& 57 & $\g \hrm^0 \to \f_k \fbar_k$ & \\
+ & 58 & $\gamma \gamma \to \f_k \fbar_k$ & \cite{Bar90} \\
& 59 & $\gamma \Z^0 \to \f_k \fbar_k$ & \\
& 60 & $\gamma \W^+ \to \f_k \fbar_l$ & \\
& 61 & $\gamma \hrm^0 \to \f_k \fbar_k$ & \\
& 62 & $\Z^0 \Z^0 \to \f_k \fbar_k$ & \\
& 63 & $\Z^0 \W^+ \to \f_k \fbar_l$ & \\
& 64 & $\Z^0 \hrm^0 \to \f_k \fbar_k$ & \\
& 65 & $\W^+ \W^- \to \f_k \fbar_k$ & \\
& 66 & $\W^+ \hrm^0 \to \f_k \fbar_l$ & \\
& 67 & $\hrm^0 \hrm^0 \to \f_k \fbar_k$ & \\
+ & 68 & $\g \g \to \g \g$ & \cite{Com77,Ben84} \\
+ & 69 & $\gamma \gamma \to \W^+ \W^-$ & \cite{Kat83} \\
+ & 70 & $\gamma \W^+ \to \Z^0 \W^+$ & \cite{Kun87} \\
+ & 71 & $\Z^0 \Z^0 \to \Z^0 \Z^0$ (longitudinal) & \cite{Abb87} \\
+ & 72 & $\Z^0 \Z^0 \to \W^+ \W^-$ (longitudinal) & \cite{Abb87} \\
+ & 73 & $\Z^0 \W^+ \to \Z^0 \W^+$ (longitudinal) & \cite{Dob91} \\
& 74 & $\Z^0 \hrm^0 \to \Z^0 \hrm^0$ & \\
& 75 & $\W^+ \W^- \to \gamma \gamma$ & \\
\hline
\end{tabular}
\end{center}
\end{table}
\begin{table}[pt]
\caption{Subprocess codes, part 3. Comments as before
\protect\label{t:procthree} }
\begin{center}
\begin{tabular}{|c|r|l|l|}
\hline
In & No. & Subprocess & Reference \\
\hline
+ & 76 & $\W^+ \W^- \to \Z^0 \Z^0$ (longitudinal) & \cite{Ben87b} \\
+ & 77 & $\W^+ \W^{\pm} \to \W^+ \W^{\pm}$ (longitudinal) &
\cite{Dun86,Bar90a} \\
& 78 & $\W^+ \hrm^0 \to \W^+ \hrm^0$ & \\
& 79 & $\hrm^0 \hrm^0 \to \hrm^0 \hrm^0$ & \\
+ & 80 & $\q_i \gamma \to \q_k \pi^{\pm}$ & \cite{Bag82} \\
+ & 81 & $\f_i \fbar_i \to \Q_k \Qbar_k$ & \cite{Com79} \\
+ & 82 & $\g \g \to \Q_k \Qbar_k$ & \cite{Com79} \\
+ & 83 & $\q_i \f_j \to \Q_k \f_l$ & \cite{Dic86} \\
+ & 84 & $\g \gamma \to \Q_k \Qbar_k$ & \cite{Fon81} \\
+ & 85 & $\gamma \gamma \to \F_k \Fbar_k$ & \cite{Bar90} \\
+ & 86 & $\g \g \to \Jpsi \g$ & \cite{Bai83} \\
+ & 87 & $\g \g \to \chi_{0 \c} \g$ & \cite{Gas87} \\
+ & 88 & $\g \g \to \chi_{1 \c} \g$ & \cite{Gas87} \\
+ & 89 & $\g \g \to \chi_{2 \c} \g$ & \cite{Gas87} \\
+ & 91 & elastic scattering & \cite{Sch94} \\
+ & 92 & single diffraction ($AB \to XB$) & \cite{Sch94} \\
+ & 93 & single diffraction ($AB \to AX$) & \cite{Sch94} \\
+ & 94 & double diffraction & \cite{Sch94} \\
+ & 95 & low-$\pT$ production & \cite{Sjo87a} \\
+ & 96 & semihard QCD $2 \to 2$ & \cite{Sjo87a} \\
+ & 99 & $\gast \q \to \q$ & \cite{Fri00} \\
& 101 & $\g \g \to \Z^0$ & \\
+ & 102 & $\g \g \to \hrm^0$ & \cite{Eic84} \\
+ & 103 & $\gamma \gamma \to \hrm^0$ & \cite{Dre89} \\
+ & 104 & $\g \g \to \chi_{0 \c}$ & \cite{Bai83} \\
+ & 105 & $\g \g \to \chi_{2 \c}$ & \cite{Bai83} \\
+ & 106 & $\g \g \to \Jpsi \gamma$ & \cite{Dre91} \\
+ & 107 & $\g \gamma \to \Jpsi \g$ & \cite{Ber81} \\
+ & 108 & $\gamma \gamma \to \Jpsi \gamma$ & \cite{Jun97} \\
+ & 110 & $\f_i \fbar_i \to \gamma \hrm^0$ & \cite{Ber85a} \\
+ & 111 & $\f_i \fbar_i \to \g \hrm^0$ & \cite{Ell88} \\
+ & 112 & $\f_i \g \to \f_i \hrm^0$ & \cite{Ell88} \\
+ & 113 & $\g \g \to \g \hrm^0$ & \cite{Ell88} \\
+ & 114 & $\g \g \to \gamma \gamma$ & \cite{Con71,Ber84,Dic88} \\
+ & 115 & $\g \g \to \g \gamma$ & \cite{Con71,Ber84,Dic88} \\
& 116 & $\g \g \to \gamma \Z^0$ & \\
& 117 & $\g \g \to \Z^0 \Z^0$ & \\
& 118 & $\g \g \to \W^+ \W^-$ & \\
\hline
\end{tabular}
\end{center}
\end{table}
\begin{table}[pt]
\caption{Subprocess codes, part 4. Comments as before.
\protect\label{t:procfour} }
\begin{center}
\begin{tabular}{|c|r|l|l|}
\hline
In & No. & Subprocess & Reference \\
\hline
& 119 & $\gamma \gamma \to \g \g$ & \\
+ & 121 & $\g \g \to \Q_k \Qbar_k \hrm^0$ & \cite{Kun84} \\
+ & 122 & $\q_i \qbar_i \to \Q_k \Qbar_k \hrm^0$ & \cite{Kun84} \\
+ & 123 & $\f_i \f_j \to \f_i \f_j \hrm^0$ ($\Z \Z$ fusion) &
\cite{Cah84} \\
+ & 124 & $\f_i \f_j \to \f_k \f_l \hrm^0$ ($\W^+\W^-$ fusion) &
\cite{Cah84} \\
+ & 131 & $\f_i \gast_{\mrm{T}} \to \f_i \g$ & \cite{Alt78} \\
+ & 132 & $\f_i \gast_{\mrm{L}} \to \f_i \g$ & \cite{Alt78} \\
+ & 133 & $\f_i \gast_{\mrm{T}} \to \f_i \gamma$ & \cite{Alt78} \\
+ & 134 & $\f_i \gast_{\mrm{L}} \to \f_i \gamma$ & \cite{Alt78} \\
+ & 135 & $\g \gast_{\mrm{T}} \to \f_i \fbar_i$ & \cite{Alt78} \\
+ & 136 & $\g \gast_{\mrm{L}} \to \f_i \fbar_i$ & \cite{Alt78} \\
+ & 137 & $\gast_{\mrm{T}} \gast_{\mrm{T}} \to \f_i \fbar_i$ &
\cite{Bai81} \\
+ & 138 & $\gast_{\mrm{T}} \gast_{\mrm{L}} \to \f_i \fbar_i$ &
\cite{Bai81} \\
+ & 139 & $\gast_{\mrm{L}} \gast_{\mrm{T}} \to \f_i \fbar_i$ &
\cite{Bai81} \\
+ & 140 & $\gast_{\mrm{L}} \gast_{\mrm{L}} \to \f_i \fbar_i$ &
\cite{Bai81} \\
+ & 141 & $\f_i \fbar_i \to \gamma/\Z^0/\Z'^0$ & \cite{Alt89} \\
+ & 142 & $\f_i \fbar_j \to \W'^+$ & \cite{Alt89} \\
+ & 143 & $\f_i \fbar_j \to \H^+$ & \cite{Gun87} \\
+ & 144 & $\f_i \fbar_j \to \R$ & \cite{Ben85a} \\
+ & 145 & $\q_i \ell_j \to \L_{\Q}$ & \cite{Wud86} \\
+ & 146 & $\e \gamma \to \e^*$ & \cite{Bau90} \\
+ & 147 & $\d \g \to \d^*$ & \cite{Bau90} \\
+ & 148 & $\u \g \to \u^*$ & \cite{Bau90} \\
+ & 149 & $\g \g \to \eta_{\mrm{tc}}$ & \cite{Eic84,App92} \\
+ & 151 & $\f_i \fbar_i \to \H^0$ & \cite{Eic84} \\
+ & 152 & $\g \g \to \H^0$ & \cite{Eic84} \\
+ & 153 & $\gamma \gamma \to \H^0$ & \cite{Dre89} \\
+ & 156 & $\f_i \fbar_i \to \A^0$ & \cite{Eic84} \\
+ & 157 & $\g \g \to \A^0$ & \cite{Eic84} \\
+ & 158 & $\gamma \gamma \to \A^0$ & \cite{Dre89} \\
+ & 161 & $\f_i \g \to \f_k \H^+$ & \cite{Bar88} \\
+ & 162 & $\q_i \g \to \ell_k \L_{\Q}$ & \cite{Hew88} \\
+ & 163 & $\g \g \to \L_{\Q} \br{\L}_{\Q}$ &
\cite{Hew88,Eic84} \\
+ & 164 & $\q_i \qbar_i \to \L_{\Q} \br{\L}_{\Q}$ &
\cite{Hew88} \\
+ & 165 & $\f_i \fbar_i \to \f_k \fbar_k$ (via $\gammaZ$) &
\cite{Eic84,Lan91} \\
+ & 166 & $\f_i \fbar_j \to \f_k \fbar_l$ (via $\W^{\pm}$) &
\cite{Eic84,Lan91} \\
+ & 167 & $\q_i \q_j \to \q_k \d^*$ & \cite{Bau90} \\
+ & 168 & $\q_i \q_j \to \q_k \u^*$ & \cite{Bau90} \\
\hline
\end{tabular}
\end{center}
\end{table}
\begin{table}[pt]
\caption{Subprocess codes, part 5. Comments as before.
\protect\label{t:procfive} }
\begin{center}
\begin{tabular}{|c|r|l|l|}
\hline
In & No. & Subprocess & Reference \\
\hline
+ & 169 & $\q_i \qbar_i \to \e^{\pm} \e^{*\mp}$ & \cite{Bau90} \\
+ & 171 & $\f_i \fbar_i \to \Z^0 \H^0$ & \cite{Eic84} \\
+ & 172 & $\f_i \fbar_j \to \W^+ \H^0$ & \cite{Eic84} \\
+ & 173 & $\f_i \f_j \to \f_i \f_j \H^0$ ($\Z \Z$ fusion) &
\cite{Cah84} \\
+ & 174 & $\f_i \f_j \to \f_k \f_l \H^0$ ($\W^+\W^-$ fusion) &
\cite{Cah84} \\
+ & 176 & $\f_i \fbar_i \to \Z^0 \A^0$ & \cite{Eic84} \\
+ & 177 & $\f_i \fbar_j \to \W^+ \A^0$ & \cite{Eic84} \\
+ & 178 & $\f_i \f_j \to \f_i \f_j \A^0$ ($\Z \Z$ fusion) &
\cite{Cah84} \\
+ & 179 & $\f_i \f_j \to \f_k \f_l \A^0$ ($\W^+\W^-$ fusion) &
\cite{Cah84} \\
+ & 181 & $\g \g \to \Q_k \Qbar_k \H^0$ & \cite{Kun84} \\
+ & 182 & $\q_i \qbar_i \to \Q_k \Qbar_k \H^0$ & \cite{Kun84} \\
+ & 183 & $\f_i \fbar_i \to \g \H^0$ & \cite{Ell88} \\
+ & 184 & $\f_i \g \to \f_i \H^0$ & \cite{Ell88} \\
+ & 185 & $\g \g \to \g \H^0$ & \cite{Ell88} \\
+ & 186 & $\g \g \to \Q_k \Qbar_k \A^0$ & \cite{Kun84} \\
+ & 187 & $\q_i \qbar_i \to \Q_k \Qbar_k \A^0$ & \cite{Kun84} \\
+ & 188 & $\f_i \fbar_i \to \g \A^0$ & \cite{Ell88} \\
+ & 189 & $\f_i \g \to \f_i \A^0$ & \cite{Ell88} \\
+ & 190 & $\g \g \to \g \A^0$ & \cite{Ell88} \\
+ & 191 & $\f_i \fbar_i \to \rho^0_{\mrm{tc}}$ & \cite{Eic96} \\
+ & 192 & $\f_i \fbar_j \to \rho^{\pm}_{\mrm{tc}}$ & \cite{Eic96} \\
+ & 193 & $\f_i \fbar_i \to \omega^0_{\mrm{tc}}$ & \cite{Eic96} \\
+ & 194 & $\f_i \fbar_i \to \f_k \fbar_k$ & \cite{Eic96,Lan99} \\
+ & 195 & $\f_i \fbar_j \to \f_k \fbar_l$ & \cite{Eic96,Lan99} \\
+ & 201 & $\f_i \fbar_i \to \se_L \se_L^*$ & \cite{Bar87,Daw85} \\
+ & 202 & $\f_i \fbar_i \to \se_R \se_R^*$ & \cite{Bar87,Daw85} \\
+ & 203 & $\f_i \fbar_i \to \se_L \se_R^*+\se_L^* \se_R$ &
\cite{Bar87} \\
+ & 204 & $\f_i \fbar_i \to \smu_L \smu_L^*$ & \cite{Bar87,Daw85} \\
+ & 205 & $\f_i \fbar_i \to \smu_R \smu_R^*$ & \cite{Bar87,Daw85} \\
+ & 206 & $\f_i \fbar_i\to\smu_L \smu_R^*+\smu_L^* \smu_R$ &
\cite{Bar87} \\
+ & 207 & $\f_i \fbar_i\to\stau_1 \stau_1^*$ & \cite{Bar87,Daw85} \\
+ & 208 & $\f_i \fbar_i\to\stau_2 \stau_2^*$ & \cite{Bar87,Daw85} \\
+ & 209 & $\f_i \fbar_i\to\stau_1
\stau_2^*+\stau_1^*\stau_2$&\cite{Bar87} \\
+ & 210 & $\f_i \fbar_j\to \sell_L {\snu}_{\ell}^*+
\sell_L^* \snu_{\ell}$&\cite{Daw85} \\
+ & 211 & $\f_i \fbar_j\to \stau_1
\snu_{\tau}^*+\stau_1^*\snu_{\tau}$ & \cite{Daw85} \\
+ & 212 & $\f_i \fbar_j\to \stau_2
\snu_{\tau}{}^*+\stau_2^*\snu_{\tau}$
& \cite{Daw85} \\
+ & 213 & $\f_i \fbar_i\to \snu_{\ell} \snu_{\ell}^*$ &
\cite{Bar87,Daw85} \\
+ & 214 & $\f_i \fbar_i\to \snu_{\tau} \snu_{\tau}^*$
& \cite{Bar87,Daw85} \\
\hline
\end{tabular}
\end{center}
\end{table}
\begin{table}[pt]
\caption{Subprocess codes, part 6. Comments as before.
\protect\label{t:procsix}}
\begin{center}
\begin{tabular}{|c|r|l|l|}
\hline
In & No. & Subprocess & Reference \\
\hline
+ & 216 & $\f_i \fbar_i \to \chio_1 \chio_1$ & \cite{Bar86a} \\
+ & 217 & $\f_i \fbar_i \to \chio_2 \chio_2$ & \cite{Bar86a} \\
+ & 218 & $\f_i \fbar_i \to \chio_3 \chio_3$ & \cite{Bar86a} \\
+ & 219 & $\f_i \fbar_i \to \chio_4 \chio_4$ & \cite{Bar86a} \\
+ & 220 & $\f_i \fbar_i \to \chio_1 \chio_2$ & \cite{Bar86a} \\
+ & 221 & $\f_i \fbar_i \to \chio_1 \chio_3$ & \cite{Bar86a} \\
+ & 222 & $\f_i \fbar_i \to \chio_1 \chio_4$ & \cite{Bar86a} \\
+ & 223 & $\f_i \fbar_i \to \chio_2 \chio_3$ & \cite{Bar86a} \\
+ & 224 & $\f_i \fbar_i \to \chio_2 \chio_4$ & \cite{Bar86a} \\
+ & 225 & $\f_i \fbar_i \to \chio_3 \chio_4$ & \cite{Bar86a} \\
+ & 226 & $\f_i \fbar_i \to \chip_1 \chim_1$ & \cite{Bar86b} \\
+ & 227 & $\f_i \fbar_i \to \chip_2 \chim_2$ & \cite{Bar86b} \\
+ & 228 & $\f_i \fbar_i \to \chip_1 \chim_2$ & \cite{Bar86b} \\
+ & 229 & $\f_i \fbar_j \to \chio_1 \chip_1$ & \cite{Bar86a,Bar86b} \\
+ & 230 & $\f_i \fbar_j \to \chio_2 \chip_1$ & \cite{Bar86a,Bar86b} \\
+ & 231 & $\f_i \fbar_j \to \chio_3 \chip_1$ & \cite{Bar86a,Bar86b} \\
+ & 232 & $\f_i \fbar_j \to \chio_4 \chip_1$ & \cite{Bar86a,Bar86b} \\
+ & 233 & $\f_i \fbar_j \to \chio_1 \chip_2$ & \cite{Bar86a,Bar86b} \\
+ & 234 & $\f_i \fbar_j \to \chio_2 \chip_2$ & \cite{Bar86a,Bar86b} \\
+ & 235 & $\f_i \fbar_j \to \chio_3 \chip_2$ & \cite{Bar86a,Bar86b} \\
+ & 236 & $\f_i \fbar_j \to \chio_4 \chip_2$ & \cite{Bar86a,Bar86b} \\
+ & 237 & $\f_i \fbar_i \to \glu \chio_1$ & \cite{Daw85} \\
+ & 238 & $\f_i \fbar_i \to \glu \chio_2$ & \cite{Daw85} \\
+ & 239 & $\f_i \fbar_i \to \glu \chio_3$ & \cite{Daw85} \\
+ & 240 & $\f_i \fbar_i \to \glu \chio_4$ & \cite{Daw85} \\
+ & 241 & $\f_i \fbar_j \to \glu \chip_1$ & \cite{Daw85} \\
+ & 242 & $\f_i \fbar_j \to \glu \chip_2$ & \cite{Daw85} \\
+ & 243 & $\f_i \fbar_i \to \glu \glu$ & \cite{Daw85} \\
+ & 244 & $\g \g \to \glu \glu$ & \cite{Daw85} \\
+ & 246 & $\f_i \g \to {\sq_i}{}_L \chio_1$ & \cite{Daw85} \\
+ & 247 & $\f_i \g \to {\sq_i}{}_R \chio_1$ & \cite{Daw85} \\
+ & 248 & $\f_i \g \to {\sq_i}{}_L \chio_2$ & \cite{Daw85} \\
+ & 249 & $\f_i \g \to {\sq_i}{}_R \chio_2$ & \cite{Daw85} \\
+ & 250 & $\f_i \g \to {\sq_i}{}_L \chio_3$ & \cite{Daw85} \\
+ & 251 & $\f_i \g \to {\sq_i}{}_R \chio_3$ & \cite{Daw85} \\
+ & 252 & $\f_i \g \to {\sq_i}{}_L \chio_4$ & \cite{Daw85} \\
+ & 253 & $\f_i \g \to {\sq_i}{}_R \chio_4$ & \cite{Daw85} \\
+ & 254 & $\f_i \g \to {\sq_j}{}_L \chip_1$ & \cite{Daw85} \\
\hline
\end{tabular}
\end{center}
\end{table}
\begin{table}[pt]
\caption{Subprocess codes, part 7. Comments as before.
\protect\label{t:procseven}}
\begin{center}
\begin{tabular}{|c|r|l|l|}
\hline
In & No. & Subprocess & Reference \\
\hline
+ & 256 & $\f_i \g \to {\sq_j}{}_L \chip_2$ & \cite{Daw85} \\
+ & 258 & $\f_i \g \to {\sq_i}{}_L \glu$ & \cite{Daw85}\\
+ & 259 & $\f_i \g \to {\sq_i}{}_R \glu$ & \cite{Daw85}\\
+ & 261 & $\f_i \fbar_i \to \tp_1 \tm_1$ & \cite{Daw85} \\
+ & 262 & $\f_i \fbar_i \to \tp_2 \tm_2$ & \cite{Daw85} \\
+ & 263 & $\f_i \fbar_i \to \tp_1 \tm_2+\tm_1 \tp_2$ & \cite{Daw85} \\
+ & 264 & $\g \g \to \tp_1 \tm_1$ & \cite{Daw85} \\
+ & 265 & $\g \g \to \tp_2 \tm_2$ & \cite{Daw85} \\
+ & 271 & $\f_i \f_j \to {\sq_i}{}_L {\sq_j}{}_L$ & \cite{Daw85} \\
+ & 272 & $\f_i \f_j \to {\sq_i}{}_R {\sq_j}{}_R$ & \cite{Daw85} \\
+ & 273 & $\f_i \f_j \to {\sq_i}{}_L {\sq_j}{}_R+
{\sq_i}{}_R {\sq_j}{}_L$ & \cite{Daw85} \\
+ & 274 & $\f_i \fbar_j \to {\sq_i}{}_L {\sqs_j}{}_L$ & \cite{Daw85} \\
+ & 275 & $\f_i \fbar_j \to {\sq_i}{}_R {\sqs_j}{}_R$ & \cite{Daw85} \\
+ & 276 & $\f_i \fbar_j \to {\sq_i}{}_L {\sqs_j}{}_R+
{\sq_i}{}_R {\sqs_j}{}_L$ & \cite{Daw85} \\
+ & 277 & $\f_i \fbar_i \to {\sq_j}{}_L {\sqs_j}{}_L$ & \cite{Daw85} \\
+ & 278 & $\f_i \fbar_i \to {\sq_j}{}_R {\sqs_j}{}_R$ & \cite{Daw85} \\
+ & 279 & $\g \g \to {\sq_i}{}_L {\sqs_i}{}_L$ & \cite{Daw85} \\
+ & 280 & $\g \g \to {\sq_i}{}_R {\sqs_i}{}_R$ & \cite{Daw85} \\
+ & 281 & $\b \q \to \sbo_1 \sq_L$ ($\q$ not $\b$) & \cite{Daw85a} \\
+ & 282 & $\b \q \to \sbo_2 \sq_R$ & \cite{Daw85a} \\
+ & 283 & $\b \q \to \sbo_1 \sq_R + \sbo2 \sq_L$ & \cite{Daw85a} \\
+ & 284 & $\b \qbar \to \sbo_1 \sqs_L$ & \cite{Daw85a} \\
+ & 285 & $\b \qbar \to \sbo_2 \sqs_R$ & \cite{Daw85a} \\
+ & 286 & $\b \qbar \to \sbo_1 \sqs_R + \sbo_2 \sqs_L$ & \cite{Daw85a} \\
+ & 287 & $\f_i \fbar_i \to \sbo_1 \sbs_1$ & \cite{Daw85a} \\
+ & 288 & $\f_i \fbar_i \to \sbo_2 \sbs_2$ & \cite{Daw85a} \\
+ & 289 & $\g \g \to \sbo_1 \sbs_1$ & \cite{Daw85a} \\
+ & 290 & $\g \g \to \sbo_2 \sbs_2$ & \cite{Daw85a} \\
+ & 291 & $\b \b \to \sbo_1 \sbo_1$ & \cite{Daw85a} \\
+ & 292 & $\b \b \to \sbo_2 \sbo_2$ & \cite{Daw85a} \\
+ & 293 & $\b \b \to \sbo_1 \sbo_2$ & \cite{Daw85a} \\
+ & 294 & $\b \g \to \sbo_1 \glu$ & \cite{Daw85a} \\
+ & 295 & $\b \g \to \sbo_2 \glu$ & \cite{Daw85a} \\
+ & 296 & $\b \bbar \to \sbo_1 \sbs_2 + \sbs_1 \sbo_2$ & \cite{Daw85a} \\
+ & 297 & $\f_i \fbar_j \to \H^{\pm} \hrm^0$ & \cite{Daw85a} \\
+ & 298 & $\f_i \fbar_j \to \H^{\pm} \H^0$ & \cite{Daw85a} \\
+ & 299 & $\f_i \fbar_i \to \A \hrm^0$ & \cite{Daw85a} \\
+ & 300 & $\f_i \fbar_i \to \A \H^0$ & \cite{Daw85a} \\
\hline
\end{tabular}
\end{center}
\end{table}
\begin{table}[pt]
\caption{Subprocess codes, part 8. Comments as before.
\protect\label{t:proceight} }
\begin{center}
\begin{tabular}{|c|r|l|l|}
\hline
In & No. & Subprocess & Reference \\
\hline
+ & 301 & $\f_i \fbar_i \to \H^+ \H^-$ & \cite{Daw85a} \\
+ & 341 & $\ell_i \ell_j \to \H_L^{\pm\pm}$ & \cite{Hui97} \\
+ & 342 & $\ell_i \ell_j \to \H_R^{\pm\pm}$ & \cite{Hui97} \\
+ & 343 & $\ell_i \gamma \to \H_L^{\pm\pm} \e^{\mp}$ & \cite{Hui97} \\
+ & 344 & $\ell_i \gamma \to \H_R^{\pm\pm} \e^{\mp}$ & \cite{Hui97} \\
+ & 345 & $\ell_i \gamma \to \H_L^{\pm\pm} \mu^{\mp}$ & \cite{Hui97} \\
+ & 346 & $\ell_i \gamma \to \H_R^{\pm\pm} \mu^{\mp}$ & \cite{Hui97} \\
+ & 347 & $\ell_i \gamma \to \H_L^{\pm\pm} \tau^{\mp}$ & \cite{Hui97} \\
+ & 348 & $\ell_i \gamma \to \H_R^{\pm\pm} \tau^{\mp}$ & \cite{Hui97} \\
+ & 349 & $\f_i \fbar_i \to \H_L^{++} \H_L^{--}$ & \cite{Hui97} \\
+ & 350 & $\f_i \fbar_i \to \H_R^{++} \H_R^{--}$ & \cite{Hui97} \\
+ & 351 & $\f_i \f_j \to \f_k f_l \H_L^{\pm\pm}$ ($\W\W$) fusion) &
\cite{Hui97} \\
+ & 352 & $\f_i \f_j \to \f_k f_l \H_R^{\pm\pm}$ ($\W\W$) fusion) &
\cite{Hui97} \\
+ & 353 & $\f_i \fbar_i \to \Z_R^0$ & \cite{Eic84} \\
+ & 354 & $\f_i \fbar_j \to \W_R^+$ & \cite{Eic84} \\
+ & 361 & $\f_i \fbar_i \to \W^+_{\mrm{L}} \W^-_{\mrm{L}} $ &
\cite{Lan99} \\
+ & 362 & $\f_i \fbar_i \to \W^{\pm}_{\mrm{L}} \pi^{\mp}_{\mrm{tc}}$ &
\cite{Lan99} \\
+ & 363 & $\f_i \fbar_i \to \pi^+_{\mrm{tc}} \pi^-_{\mrm{tc}}$ &
\cite{Lan99} \\
+ & 364 & $\f_i \fbar_i \to \gamma \pi^0_{\mrm{tc}} $ & \cite{Lan99} \\
+ & 365 & $\f_i \fbar_i \to \gamma {\pi'}^0_{\mrm{tc}} $ & \cite{Lan99} \\
+ & 366 & $\f_i \fbar_i \to \Z^0 \pi^0_{\mrm{tc}} $ & \cite{Lan99} \\
+ & 367 & $\f_i \fbar_i \to \Z^0 {\pi'}^0_{\mrm{tc}} $ & \cite{Lan99} \\
+ & 368 & $\f_i \fbar_i \to \W^{\pm} \pi^{\mp}_{\mrm{tc}}$ & \cite{Lan99} \\
+ & 370 & $\f_i \fbar_j \to \W^{\pm}_{\mrm{L}} \Z^0_{\mrm{L}}$ &
\cite{Lan99} \\
+ & 371 & $\f_i \fbar_j \to \W^{\pm}_{\mrm{L}} \pi^0_{\mrm{tc}}$ &
\cite{Lan99} \\
+ & 372 & $\f_i \fbar_j \to \pi^{\pm}_{\mrm{tc}} \Z^0_{\mrm{L}} $ &
\cite{Lan99} \\
+ & 373 & $\f_i \fbar_j \to \pi^{\pm}_{\mrm{tc}} \pi^0_{\mrm{tc}} $ &
\cite{Lan99} \\
+ & 374 & $\f_i \fbar_j \to \gamma \pi^{\pm}_{\mrm{tc}} $ & \cite{Lan99} \\
+ & 375 & $\f_i \fbar_j \to \Z^0 \pi^{\pm}_{\mrm{tc}} $ & \cite{Lan99} \\
+ & 376 & $\f_i \fbar_j \to \W^{\pm} \pi^0_{\mrm{tc}} $ & \cite{Lan99} \\
+ & 377 & $\f_i \fbar_j \to \W^{\pm} {\pi'}^0_{\mrm{tc}}$ & \cite{Lan99} \\
+ & 381 & $\q_i \q_j \to \q_i \q_j$ (QCD+TC) & \cite{Chi90,Lan02a} \\
+ & 382 & $\q_i \qbar_i \to \q_k \qbar_k$ (QCD+TC) & \cite{Chi90,Lan02a} \\
+ & 383 & $\q_i \qbar_i \to \g \g$ (QCD+TC) & \cite{Lan02a} \\
+ & 384 & $\f_i \g \to \f_i \g$ (QCD+TC) & \cite{Lan02a} \\
+ & 385 & $\g \g \to \q_k \qbar_k$ (QCD+TC) & \cite{Lan02a} \\
+ & 386 & $\g \g \to \g \g$ (QCD+TC) & \cite{Lan02a} \\
+ & 387 & $\f_i \fbar_i \to \Q_k \Qbar_k$ (QCD+TC) & \cite{Lan02a} \\
\hline
\end{tabular}
\end{center}
\end{table}
\begin{table}[pt]
\caption{Subprocess codes, part 9. Comments as before.
\protect\label{t:procnine} }
\begin{center}
\begin{tabular}{|c|r|l|l|}
\hline
In & No. & Subprocess & Reference \\
\hline
+ & 388 & $\g \g \to \Q_k \Qbar_k$ (QCD+TC) & \cite{Lan02a} \\
+ & 391 & $\f \fbar \to \G^*$ & \cite{Ran99} \\
+ & 392 & $\g \g \to \G^*$ & \cite{Ran99} \\
+ & 393 & $\q \qbar \to \g \G^*$ & \cite{Ran99,Bij01} \\
+ & 394 & $\q \g \to \q \G^*$ & \cite{Ran99,Bij01} \\
+ & 395 & $\g \g \to \g \G^*$ & \cite{Ran99,Bij01} \\
+ & 401 & $\g \g \to \tbar \b \H^+$ & \cite{Bor99} \\
+ & 402 & $\q \qbar \to \tbar \b \H^+$ & \cite{Bor99} \\
+ & 421 & $\g \g \to \c\cbar[^3S_1^{(1)}] \, \g$ & \cite{Bod95} \\
+ & 422 & $\g \g \to \c\cbar[^3S_1^{(8)}] \, \g$ & \cite{Bod95} \\
+ & 423 & $\g \g \to \c\cbar[^1S_0^{(8)}] \, \g$ & \cite{Bod95} \\
+ & 424 & $\g \g \to \c\cbar[^3P_J^{(8)}] \, \g$ & \cite{Bod95} \\
+ & 425 & $\g \q \to \q \, \c\cbar[^3S_1^{(8)}]$ & \cite{Bod95} \\
+ & 426 & $\g \q \to \q \, \c\cbar[^1S_0^{(8)}]$ & \cite{Bod95} \\
+ & 427 & $\g \q \to \q \, \c\cbar[^3P_J^{(8)}]$ & \cite{Bod95} \\
+ & 428 & $\q \qbar \to \g \, \c\cbar[^3S_1^{(8)}]$ & \cite{Bod95} \\
+ & 429 & $\q \qbar \to \g \, \c\cbar[^1S_0^{(8)}]$ & \cite{Bod95} \\
+ & 430 & $\q \qbar \to \g \, \c\cbar[^3P_J^{(8)}]$ & \cite{Bod95} \\
+ & 431 & $\g \g \to \c\cbar[^3P_0^{(1)}] \, \g$ & \cite{Bod95} \\
+ & 432 & $\g \g \to \c\cbar[^3P_1^{(1)}] \, \g$ & \cite{Bod95} \\
+ & 433 & $\g \g \to \c\cbar[^3P_2^{(1)}] \, \g$ & \cite{Bod95} \\
+ & 434 & $\g \q \to \q \, \c\cbar[^3P_0^{(1)}]$ & \cite{Bod95} \\
+ & 435 & $\g \q \to \q \, \c\cbar[^3P_1^{(1)}]$ & \cite{Bod95} \\
+ & 436 & $\g \q \to \q \, \c\cbar[^3P_2^{(1)}]$ & \cite{Bod95} \\
+ & 437 & $\q \qbar \to \g \, \c\cbar[^3P_0^{(1)}]$ & \cite{Bod95} \\
+ & 438 & $\q \qbar \to \g \, \c\cbar[^3P_1^{(1)}]$ & \cite{Bod95} \\
+ & 439 & $\q \qbar \to \g \, \c\cbar[^3P_2^{(1)}]$ & \cite{Bod95} \\
+ & 461 & $\g \g \to \b\bbar[^3S_1^{(1)}] \, \g$ & \cite{Bod95} \\
+ & 462 & $\g \g \to \b\bbar[^3S_1^{(8)}] \, \g$ & \cite{Bod95} \\
+ & 463 & $\g \g \to \b\bbar[^1S_0^{(8)}] \, \g$ & \cite{Bod95} \\
+ & 464 & $\g \g \to \b\bbar[^3P_J^{(8)}] \, \g$ & \cite{Bod95} \\
+ & 465 & $\g \q \to \q \, \b\bbar[^3S_1^{(8)}]$ & \cite{Bod95} \\
+ & 466 & $\g \q \to \q \, \b\bbar[^1S_0^{(8)}]$ & \cite{Bod95} \\
+ & 467 & $\g \q \to \q \, \b\bbar[^3P_J^{(8)}]$ & \cite{Bod95} \\
+ & 468 & $\q \qbar \to \g \, \b\bbar[^3S_1^{(8)}]$ & \cite{Bod95} \\
+ & 469 & $\q \qbar \to \g \, \b\bbar[^1S_0^{(8)}]$ & \cite{Bod95} \\
\hline
\end{tabular}
\end{center}
\end{table}
\clearpage
\begin{table}[pt]
\caption{Subprocess codes, part 10. Comments as before.
\protect\label{t:procten} }
\begin{center}
\begin{tabular}{|c|r|l|l|}
\hline
In & No. & Subprocess & Reference \\
\hline
+ & 470 & $\q \qbar \to \g \, \b\bbar[^3P_J^{(8)}]$ & \cite{Bod95} \\
+ & 471 & $\g \g \to \b\bbar[^3P_0^{(1)}] \, \g$ & \cite{Bod95} \\
+ & 472 & $\g \g \to \b\bbar[^3P_1^{(1)}] \, \g$ & \cite{Bod95} \\
+ & 473 & $\g \g \to \b\bbar[^3P_2^{(1)}] \, \g$ & \cite{Bod95} \\
+ & 474 & $\g \q \to \q \, \b\bbar[^3P_0^{(1)}]$ & \cite{Bod95} \\
+ & 475 & $\g \q \to \q \, \b\bbar[^3P_1^{(1)}]$ & \cite{Bod95} \\
+ & 476 & $\g \q \to \q \, \b\bbar[^3P_2^{(1)}]$ & \cite{Bod95} \\
+ & 477 & $\q \qbar \to \g \, \b\bbar[^3P_0^{(1)}]$ & \cite{Bod95} \\
+ & 478 & $\q \qbar \to \g \, \b\bbar[^3P_1^{(1)}]$ & \cite{Bod95} \\
+ & 479 & $\q \qbar \to \g \, \b\bbar[^3P_2^{(1)}]$ & \cite{Bod95} \\
\hline
\end{tabular}
\end{center}
\end{table}
%referee: section title is missing
\section{Physics Processes}
\label{s:pytproc}
In this section we enumerate the physics processes that are available
in {\Py}, introducing the {\ISUB} code that can be used to select
desired processes.
A number of comments are made about the
physics scenarios involved, with emphasis on the
underlying assumptions and domain of validity. The section closes
with a survey of interesting processes by machine.
Note that {\ISUB} is a dummy index introduced
to allow simple referencing of processes. There are no global
variables in {\Py} named {\ISUB}.
\subsection{The Process Classification Scheme}
\label{ss:ISUBcode}
A wide selection of fundamental $2 \to 1$ and $2 \to 2$ tree
processes of the Standard Model (electroweak and strong) has been
included in {\Py}, and slots are provided for some more, not (yet)
implemented. In addition, `minimum-bias'-type processes
(like elastic scattering), loop graphs, box graphs, $2 \to 3$ tree
graphs and many non-Standard Model processes are included. The
classification is not always unique. A process that proceeds only
via an $s$-channel state is classified as a $2 \to 1$ process
(e.g.\ $\q \qbar \to \gammaZ \to \ee$).
A generic $2 \to 2$ process
may have contributions from $s$-, $t-$ and $u$-channel diagrams.
Also, in the program, $2 \to 1$ and
$2 \to 2$ graphs may sometimes be convoluted with two $1 \to 2$
splittings to form effective $2 \to 3$ or $2 \to 4$ processes
($\W^+ \W^- \to \hrm^0$ is folded with $\q \to \q'' \W^+$ and
$\q' \to \q''' \W^-$ to give $\q \q' \to \q'' \q''' \hrm^0$).
The original classification and numbering scheme is less relevant
today than when originally conceived. The calculation
of $2 \to 3$ or $2 \to 4$ matrix elements by hand is sufficiently complicated
that approximation schemes were employed, such as the effective
$\W$-approximation which factored $\W$ bosons into an effective parton
density.
Today, improvements in computational
techniques and increases in computing power make the exact calculation
manageable.
Given the
large top mass and large Higgs boson mass limits, there is also a natural
subdivision, such that the $\b$ quark is the heaviest object for which
the parton-distribution concept makes sense at current or near-future
colliders. Therefore most of the
prepared but empty slots are likely to remain empty, or be reclaimed
for other processes.
It is possible to select a combination of subprocesses
and also to know which subprocess was actually selected in
each event. For this purpose, all subprocesses are numbered according
to an {\ISUB} code. The list of possible codes is given in Tables
%\ref{t:procone} through \ref{t:procten}.
%referee: mention Appendix A in the text
\ref{t:procone} through \ref{t:procten}, and summarized in Appendix A.
Only processes marked with a `+' sign in the first
column have been implemented in the program to date. Although
{\ISUB} codes were originally designed in a logical fashion,
subsequent developments of the program have
obscured the structure. For instance, the process numbers for Higgs
production are spread out, in part as a consequence of the original
classification, in part because further production mechanisms have been
added one at a time, in whatever free slots could be found.
In the thematic descriptions that follow the main tables, the
processes of interest are repeated in a more logical order. If you
want to look for a specific process, it will be easier
to find it there.
In the following, $\f_i$ represents a fundamental fermion of flavour
$i$, i.e.\ $\d$, $\u$, $\s$, $\c$, $\b$, $\t$, $\b'$,
$\t'$, $\e^-$, $\nu_{\e}$, $\mu^-$, $\nu_{\mu}$, $\tau^-$,
$\nu_{\tau}$, ${\tau'}^-$ or ${\nu'}_{\tau}$. A corresponding antifermion
is denoted by $\fbar_i$. In several cases, some classes of fermions
are explicitly excluded, since they do not couple to the $\g$ or
$\gamma$ (no $\ee \to \g \g$, e.g.). When processes have only been
included for quarks, while leptons might also have been possible,
the notation $\q_i$ is used. A lepton is denoted by $\ell$; in a few
cases neutrinos are also lumped under this heading.
In processes where fermion masses are explicitly included in the
matrix elements, an $\F$ or $\Q$ is used to denote
an arbitrary fermion or quark. Flavours appearing already in
the initial state are denoted by indices $i$ and $j$, whereas new
flavours in the final state are denoted by $k$ and~$l$.
In supersymmetric processes, antiparticles of sfermions
are denoted by $^*$, i.e.\ $\st^*$.
Charge-conjugate channels are always assumed included as well (where
separate), and processes involving a $\W^+$ also imply those involving
a $\W^-$. Wherever $\Z^0$ is written, it is understood that $\gamma^*$
and $\gammaZ$ interference should be included as well (with
possibilities to switch off either, if so desired).
In practice, this means that fermion pairs produced from $\gammaZ$ decay
will have invariant masses as small as the program cutoff, and not
regulated by the large $\Z$ mass. The cutoff is set by an appropriate
$\ttt{CKIN}$ variable. In some cases, $\gammaZ$ interference
is not implemented; see further below.
Correspondingly, $\Z'^0$ denotes the complete set
$\gamma^*/\Z^0/\Z'^0$ (or some subset of it). Thus the notation
$\gamma$ is only used for a photon on the mass shell.
In the last column of the tables below, references are given to works
from which formulae have been taken. Sometimes these references are to
the original works on the subject, sometimes only to the place where
the formulae are given in the most convenient or accessible form, or
where chance lead us. Apologies to all matrix-element calculators who
are not mentioned. However, remember that this is not a review article
on physics processes, but only a way for readers to know what is
actually found in the program, for better or worse. In several
instances, errata have been obtained from the authors. Often
the formulae given in the literature have been generalized to include
trivial radiative corrections, Breit--Wigner line shapes with
$\hat{s}$-dependent widths (see section \ref{ss:kinemreson}), etc.
The following sections contain some useful comments on the processes
included in the program, grouped by physics interest rather than
sequentially by {\ISUB} or \ttt{MSEL} code (see \ref{ss:PYswitchkin}
for further information on the \ttt{MSEL} code). The different
{\ISUB} and \ttt{MSEL} codes
that can be used to simulate the different groups are given. {\ISUB}
codes within brackets indicate the kind of processes that indirectly
involve the given physics topic, although only as part of a larger
whole. Some obvious examples, such as the possibility to produce jets
in just about any process, are not spelled out in detail.
The text at times contains information on which special switches or
parameters are of particular interest to a given process. All these
switches are described in detail in sections \ref{ss:PYswitchpar}
\ref{ss:coupcons} and \ref{ss:susycode}, but
are alluded to here so as to provide a more complete picture of the
possibilities available for the different subprocesses. However, the
list of possibilities is certainly not exhausted by the text below.
\subsection{QCD Processes}
Obviously most processes in {\Py} contain QCD physics one way or
another, so the above title should not be overstressed. One example:
a process like $\ee \to \gammaZ \to \q\qbar$ is also traditionally
called a QCD event, but is here book-kept as $\gammaZ$ production.
In this section we discuss scatterings between coloured partons,
plus a few processes that are close relatives to other processes
of this kind.
\subsubsection{QCD jets}
\label{sss:QCDjetclass}
\ttt{MSEL} = 1, 2 \\
{\ISUB} =
\begin{tabular}[t]{rl}
11 & $\q_i \q_j \to \q_i \q_j$ \\
12 & $\q_i \qbar_i \to \q_k \qbar_k$ \\
13 & $\q_i \qbar_i \to \g \g$ \\
28 & $\q_i \g \to \q_i \g$ \\
53 & $\g \g \to \q_k \qbar_k$ \\
68 & $\g \g \to \g \g$ \\
96 & semihard QCD $2 \to 2$ \\
\end{tabular}
These are all tree-level, $2 \to 2$ process with a cross section
$\propto \alpha_s^2$.
No higher-order loop corrections are explicitly included.
However, initial- and final-state QCD radiation added to the
above processes generates multijet events.
In general, the rate of multijet ($>2$) production through the
parton shower mechanism is less certain for jets at high-$\pT$
and for well-separated pairs of jets.
A string-based fragmentation scheme such as the Lund model needs
cross sections for the different colour flows. The planar
approximation allows such a subdivision, with cross sections,
calculated in \cite{Ben84}, that differ from the usual ones by
interference terms of the order $1/N_C^2$. By default, the standard
colour-summed QCD expressions for the differential cross sections
are used with the interference terms distributed among
the various colour flows according to the pole structure of the terms.
However, the interference terms can be excluded by changing
\ttt{MSTP(34)}.
As an example, consider subprocess 28, $\q \g \to \q \g$. The
differential
cross section for this process, obtained by summing and squaring the
Feynman graphs and employing the identity of the Mandelstam variables
for the massless case, $\hat{s} + \hat{t} + \hat{u} = 0$, is
proportional to \cite{Com77}
\begin{equation}
\frac{\hat{s}^2 + \hat{u}^2}{\hat{t}^2} -
\frac{4}{9} \left( \frac{\hat{s}}{\hat{u}} +
\frac{\hat{u}}{\hat{s}} \right) ~.
\end{equation}
On the other hand, the cross sections for the two possible colour
flows of this subprocess are \cite{Ben84}
\begin{eqnarray}
A: & & \frac{4}{9} \left( 2 \frac{\hat{u}^2}{\hat{t}^2} -
\frac{\hat{u}}{\hat{s}} \right) ~; \nonumber \\
B: & & \frac{4}{9} \left( 2 \frac{\hat{s}^2}{\hat{t}^2} -
\frac{\hat{s}}{\hat{u}} \right) ~.
\end{eqnarray}
Colour configuration $A$ is one in which the original colour
of the $\q$ annihilates with the anticolour of the $\g$,
the $\g$ colour flows through, and a new colour--anticolour is
created between the final $\q$ and $\g$. In colour configuration
$B$, the gluon anticolour flows through, but the $\q$ and $\g$
colours are interchanged. Note that these two colour
configurations have different kinematics dependence.
In principle, this has observable consequences.
For \ttt{MSTP(34) = 0}, these are the cross sections actually
used.
The sum of the $A$ and $B$ contributions is
\begin{equation}
\frac{8}{9} \frac{\hat{s}^2 + \hat{u}^2}{\hat{t}^2} -
\frac{4}{9} \left( \frac{\hat{s}}{\hat{u}} +
\frac{\hat{u}}{\hat{s}} \right) ~.
\end{equation}
The difference between this expression and that of \cite{Com77},
corresponding to the interference between the two colour-flow
configurations, is then
\begin{equation}
\frac{1}{9} \frac{\hat{s}^2 + \hat{u}^2}{\hat{t}^2} ~,
\end{equation}
which can be naturally divided between colour flows $A$ and $B$:
\begin{eqnarray}
A: & & \frac{1}{9} \frac{\hat{u}^2}{\hat{t}^2} ~; \nonumber \\
B: & & \frac{1}{9} \frac{\hat{s}^2}{\hat{t}^2} ~.
\end{eqnarray}
For \ttt{MSTP(34) = 1}, the standard QCD matrix element is
used, with the same relative importance of the two colour configurations
as above. Similar procedures are followed also for the other QCD
subprocesses.
All the matrix elements in this group are for massless quarks
(although final-state quarks are of course put on the mass shell).
As a consequence, cross sections are divergent for $\pT \to 0$,
and some kind of regularization is required. Normally you
are expected to set the desired $\pTmin$ value in
\ttt{CKIN(3)}.
The new flavour produced in the annihilation processes ({\ISUB} = 12 and
53) is determined by the flavours allowed for gluon splitting into
quark--antiquark; see switch \ttt{MDME}.
Subprocess 96 is special among all the ones in the program. In terms of
the basic cross section, it is equivalent to the sum of the other ones,
i.e.\ 11, 12, 13, 28, 53 and 68. The phase space is mapped differently,
however, and allows $\pT$ as the input variable. This is especially useful
in the context of the multiple interactions machinery (see section
\ref{ss:multint}) where potential scatterings are considered in order
of decreasing $\pT$, with a form factor related to the probability of not
having another scattering with a $\pT$ larger than the considered one.
You are not expected to access process 96 yourself. Instead it is
automatically initialized and used either if process 95 is included or
if multiple interactions are switched on. The process will then appear
in the maximization information output, but not in the cross section
table at the end of a run. Instead, the hardest scattering generated within
the context of process 95 is reclassified as an event of the 11, 12, 13,
28, 53 or 68 kinds, based on the relative cross section for these in
the point chosen. Further multiple interactions, subsequent to the hardest
one, also do not show up in cross section tables.
\subsubsection{Heavy flavours}
\label{sss:heavflavclass}
\ttt{MSEL} = 4, 5, 6, 7, 8 \\
{\ISUB} =
\begin{tabular}[t]{rl}
81 & $\q_i \qbar_i \to \Q_k \Qbar_k$ \\
82 & $\g \g \to \Q_k \Qbar_k$ \\
(83) & $\q_i \f_j \to Q_k f_l$ \\
(84) & $\g \gamma \to \Q_k \Qbar_k$ \\
(85) & $\gamma \gamma \to \F_k \Fbar_k$ \\
(1) & $\f_i \fbar_i \to \gammaZ \to \F_k \Fbar_k$ \\
(2) & $\f_i \fbar_j \to \W^+\to \F_k \Fbar_l $ \\
(142) & $\f_i \fbar_j \to \W'^+\to \F_k \Fbar_l $ \\
\end{tabular}
The matrix elements in this group differ from the corresponding ones
in the group above in that they correctly take into account the quark
masses. As a consequence, the cross sections are finite for
$\pT \to 0$ and require no special
cuts.
The two first processes that appear here are the dominant lowest-order
QCD graphs in hadron colliders --- a few other graphs will be mentioned
later, such as process 83.
The flavour produced is selected according to a hierarchy of options:
\begin{Enumerate}
\item if \ttt{MSEL = 4 - 8} then the flavour is set by the \ttt{MSEL} value;
\item else if \ttt{MSTP(7) = 1 - 8} then the flavour is set by the
\ttt{MSTP(7)} value;
\item else the flavour is determined by the heaviest flavour allowed for
gluon splitting into quark--antiquark; see switch \ttt{MDME}.
\end{Enumerate}
Note that only one heavy flavour is allowed at a time; if more than one
is turned on in \ttt{MDME}, only the heaviest will be produced (as
opposed to the case for {\ISUB} = 12 and 53 above, where more than one
flavour is allowed simultaneously).
The lowest-order processes listed above just represent one source of
heavy-flavour production. Heavy quarks can also be present in the
parton distributions at the $Q^2$ scale of the hard interaction,
leading to processes like $\Q \g \to \Q \g$,
so-called flavour excitation, or they can be
created by gluon splittings $\g \to \Q \Qbar$ in initial- or final-state
shower evolution. The implementation and importance of these various
production mechanisms is discussed in detail in \cite{Nor98}.
In fact, as the c.m.\ energy is increased, these other processes gain
in importance relative to the lowest-order production graphs above.
As as example, only 10\%--20\% of the $\b$ production at LHC
energies come from the lowest-order graphs. The figure is even smaller
for charm, while it is well above 50\% for top. At LHC energies,
the specialized treatment described in this section is therefore
only of interest for top (and potential fourth-generation quarks) ---
the higher-order corrections can here be approximated by an effective
$K$ factor, except possibly in some rare corners of phase space.
For charm and bottom, on the other hand, it is necessary to simulate the
full event sample (within the desired kinematics cuts), and then only
keep those events that contain $\b/\c$, be that either from lowest-order
production, or flavour excitation, or gluon splitting. Obviously this may
be a time-consuming enterprise --- although the probability for a high-$\pT$
event at collider energies to contain (at least) one charm or bottom pair
is fairly large, most of these heavy flavours are carrying a small fraction
of the total $\pT$ flow of the jets, and therefore do not survive normal
experimental cuts. We note that the lowest-order production of charm or
bottom with processes 12 and 53, as part of the standard QCD mix, is now
basically equivalent with that offered by processes 81 and 82. For 12
vs.\ 81 this is rather trivial, since only $s$-channel gluon exchange is
involved, but for process 53 it requires a separate evaluation of massive
matrix elements for $\c$ and $\b$ in the flavour sum.
This is performed
by retaining the $\hat{s}$ and $\hat{\theta}$ values already preliminarily
selected for the massless kinematics, and recalculating $\hat{t}$ and
$\hat{u}$ with mass effects included. Some of the documentation information
in \ttt{PARI} does not properly reflect this recalculation, but that is
purely a documentation issue. Also process 96, used internally for the
total QCD jet cross section, includes $\c$ and $\b$ masses. Only the hardest
interaction in a multiple interactions scenario may contain $\c/\b$,
however, for technical reasons, so that the total rate may be underestimated.
(Quite apart from other uncertainties, of course.)
As an aside, it is not only for the lowest-order graphs that events
may be generated with a guaranteed heavy-flavour content. One may also
generate the flavour excitation process by itself, in the massless
approximation, using {\ISUB} = 28 and setting the \ttt{KFIN} array
appropriately. No trick exists to force the gluon splittings without
introducing undesirable biases, however. In order to have it all, one
therefore has to make a full QCD jets run, as already noted.
Also other processes can generate heavy flavours, all the way up to top,
but then without a proper account of masses. By default, top production
is switched off in those processes where a new flavour pair is produced
at a gluon or photon vertex, i.e.\ 12, 53, 54, 58, 96 and 135--140, while
charm and bottom is allowed. These defaults can be changed by setting
the \ttt{MDME(IDC,1)} values of the appropriate $\g$ or $\gamma$
`decay channels'; see further below.
The cross section for heavy quark pair production close to threshold can be
modified according to the formulae of \cite{Fad90}; see \ttt{MSTP(35)}.
Here threshold effects due to $\Q\Qbar$ bound-state formation are taken
into account in a smeared-out, average sense. Then the na\"{\i}ve
cross section is multiplied by the squared wave function at the origin.
In a colour-singlet channel this gives a net enhancement of the form
\begin{equation}
|\Psi^{(s)}(0)|^2 = \frac{X_{(s)}}{1 - \exp(- X_{(s)})} ~ ,
~~~ \mrm{where}~ X_{(s)} = \frac{4}{3}
\frac{\pi \alphas}{\beta} ~,
\label{pp:threshenh}
\end{equation}
where $\beta$ is the quark velocity,
while in a colour octet channel there is a net suppression given by
\begin{equation}
%|\Psi^{(8)}(0)|^2 = \frac{X_{(8)}}{\exp(- X_{(8)}) -1} ~ ,
%referee: wrong sign
|\Psi^{(8)}(0)|^2 = \frac{X_{(8)}}{\exp(X_{(8)}) -1} ~ ,
~~~ \mrm{where}~ X_{(8)} = \frac{1}{6}
\frac{\pi \alphas}{\beta} ~.
\label{pp:threshsup}
\end{equation}
The $\alphas$ factor in this expression is related to the energy
scale of bound-state formation and is selected independently from
the factor in the standard production cross section.
The presence of a threshold factor affects the total rate and also
kinematic distributions.
Heavy flavours can also be produced by secondary decays of gauge
bosons or new exotic particles. We have listed 1, 2 and 142 above
as among the most important ones. There is a special point to
including $\W'$ in this list. Imagine that you want to study
the electroweak $s$-channel production of a single top,
$\u \dbar \to \W^+ \to \t \bbar$, and therefore decide to force this
particular decay mode of the $\W$. But then the same decay channel
is required for the $\W^+$ produced in the decay $\t \to \b \W^+$,
i.e.\ you have set up an infinite recursion
$\W \to \t \to \W \to \t \to \ldots$. The way out is to use the
$\W'$, which has default couplings just like the normal $\W$,
only a different mass, which then can be changed to agree,
\ttt{PMAS(34,1) = PMAS(24,1)}.
The $\W'$ is now forced to decay to $\t \bbar$, while the $\W$
can decay freely (or also be forced, e.g. to have a leptonic
decay, if desired). (Additionally, it may be necessary to raise
\ttt{CKIN(1)} to be at least around the top mass, so that the
program does not get stuck in a region of phase space where
the cross section is vanishing.) Alternatively, a full run
(after raising \ttt{CKIN(1)} to be just below the single top
threshold) can be used if one is willing to select the desired
events by hand.
Heavy flavours, i.e.\ top and fourth generation, are assumed to be so
short-lived that they decay before they have time to hadronize. This
means that the light quark in the decay $\Q \to \W^{\pm} \q$
inherits the colour of the heavy one. The current {\Py} description
represents a change of philosophy compared to much earlier versions,
formulated at a time when the top was thought to live long enough
to form hadrons. For event shapes the difference between the
two time orderings normally has only marginal effects \cite{Sjo92a}.
In practical terms,
the top (or a fourth generation fermion) is
treated like a resonance in the sense
of section \ref{sss:resdecaycross}, i.e.\ the cross-section is reduced
so as only to correspond to the channels left open by you.
This also includes the restrictions on secondary decays, i.e.\ on the
decays of a $\W^+$ or a $\H^+$ produced in the top decay.
For $\b$ and $\c$ quarks, which are long-lived enough to form hadrons,
no such reduction takes place.
Branching ratios then have to be folded in by hand to get the correct
cross sections.
%The logic behind this difference is that when
%hadronization takes place, one would normally decay the
%$\D^0$ and $\D^+$ meson according to different branching ratios.
%But which $\D$ mesons are to be formed is not known at the bottom quark
%creation, so one could not weight for that. For a $\t$ quark, which
%decays rapidly, this ambiguity does not exist, and so a reduction
%factor can be introduced directly coupled to the $\t$ quark production
%process.
This rule about cross-section calculations applies to all the
processes explicitly set up to handle heavy flavour creation.
In addition to the ones above, this means all the ones in Tables
\ref{t:procone}--\ref{t:proceight} where the fermion final state is
given as capital letters (`$\Q$' and `$\F$') and also flavours produced
in resonance decays ($\Z^0$, $\W^{\pm}$, $\hrm^0$, etc., including
processes 165 and 166). However, heavy flavours could also be produced
in a process such as 31, $\q_i \g \to \q_k \W^{\pm}$, where $\q_k$
could be a top quark. In this case, the thrust of the description is
clearly on light flavours --- the kinematics of the process is
formulated in the massless fermion limit --- so any top production
is purely incidental. Since here the choice of scattered flavour is
only done at a later stage, the top branching ratios are not
correctly folded in to the hard-scattering cross section. So, for
applications like these, it is not recommended to restrict the allowed
top decay modes. Often one might like to get rid of the possibility
of producing top together with light flavours. This can be
done by switching off (i.e.\ setting \ttt{MDME(I,1) = 0}) the
`channels' $\d \to \W^- \t$, $\s \to \W^- \t$, $\b \to \W^- \t$,
$\g \to \t\tbar$ and $\gamma \to \t\tbar$. Also any heavy flavours
produced by parton-shower evolution would not be correctly weighted
into the cross section. However, currently top production is switched
off both as a beam remnant (see \ttt{MSTP(9)}) and in initial
(see \ttt{KFIN} array) and final (see \ttt{MSTJ(45)}) state radiation.
In pair production of heavy flavour (top) in processes 81, 82, 84
and 85, matrix elements are only given for one common mass, although
Breit--Wigner distributions are used to select two separate masses.
As described in section \ref{ss:kinemreson}, an average mass value
is constructed for the matrix element evaluation so that the
$\beta_{34}$ kinematics factor can be retained.
Because of its large mass, it is possible that the top quark can decay
to some not yet discovered particle. Some such alternatives are included
in the program, such as $\t \to \b \H^+$ or $\t \to \grav \st$. These
decays are not obtained by default, but can be included as discussed
for the respective physics scenario.
\subsubsection{J/$\psi$ and other Hidden Heavy Flavours}
\label{sss:Jpsiclass}
{\ISUB} =
\begin{tabular}[t]{rl}
86 & $\g \g \to \Jpsi \g$ \\
87 & $\g \g \to \chi_{0 \c} \g$ \\
88 & $\g \g \to \chi_{1 \c} \g$ \\
89 & $\g \g \to \chi_{2 \c} \g$ \\
104 & $\g \g \to \chi_{0 \c}$ \\
105 & $\g \g \to \chi_{2 \c}$ \\
106 & $\g \g \to \Jpsi \gamma$ \\
107 & $\g \gamma \to \Jpsi \g$ \\
108 & $\gamma \gamma \to \Jpsi \gamma$ \\
\end{tabular}\\[2mm]
\ttt{MSEL} = 61,62,63 \\
{\ISUB} =
\begin{tabular}[t]{rrl}
$\c\cbar$ & $\b\bbar$ & \\
421 & 461 & $\g \g \to \Q\Qbar[^3S_1^{(1)}] \, \g$ \\
422 & 462 & $\g \g \to \Q\Qbar[^3S_1^{(8)}] \, \g$ \\
423 & 463 & $\g \g \to \Q\Qbar[^1S_0^{(8)}] \, \g$ \\
424 & 464 & $\g \g \to \Q\Qbar[^3P_J^{(8)}] \, \g$ \\
425 & 465 & $\g \q \to \q \, \Q\Qbar[^3S_1^{(8)}]$ \\
426 & 466 & $\g \q \to \q \, \Q\Qbar[^1S_0^{(8)}]$ \\
427 & 467 & $\g \q \to \q \, \Q\Qbar[^3P_J^{(8)}]$ \\
428 & 468 & $\q \qbar \to \g \, \Q\Qbar[^3S_1^{(8)}]$ \\
429 & 469 & $\q \qbar \to \g \, \Q\Qbar[^1S_0^{(8)}]$ \\
430 & 470 & $\q \qbar \to \g \, \Q\Qbar[^3P_J^{(8)}]$ \\
431 & 471 & $\g \g \to \Q\Qbar[^3P_0^{(1)}] \, \g$ \\
432 & 472 & $\g \g \to \Q\Qbar[^3P_1^{(1)}] \, \g$ \\
433 & 473 & $\g \g \to \Q\Qbar[^3P_2^{(1)}] \, \g$ \\
434 & 474 & $\g \q \to \q \, \Q\Qbar[^3P_0^{(1)}]$ \\
435 & 475 & $\g \q \to \q \, \Q\Qbar[^3P_1^{(1)}]$ \\
436 & 476 & $\g \q \to \q \, \Q\Qbar[^3P_2^{(1)}]$ \\
437 & 477 & $\q \qbar \to \g \, \Q\Qbar[^3P_0^{(1)}]$ \\
438 & 478 & $\q \qbar \to \g \, \Q\Qbar[^3P_1^{(1)}]$ \\
439 & 479 & $\q \qbar \to \g \, \Q\Qbar[^3P_2^{(1)}]$ \\
\end{tabular}
In {\Py} one may distinguish between three main sources of $\Jpsi$
production.
\begin{Enumerate}
\item Decays of $\B$ mesons and baryons.
\item Parton-shower evolution, wherein a $\c$ and a $\cbar$ quark
produced in adjacent branchings
(e.g.\ $\g \to \g \g \to \c \cbar \c \cbar$)
turn out to have so small an invariant mass that the pair collapses
to a single particle.
\item Direct production, where a $\c$ quark loop gives a coupling
between a set of gluons and a $\c\cbar$ bound state. Higher-lying
states, like the $\chi_c$ ones, may subsequently decay to $\Jpsi$.
\end{Enumerate}
The first two sources are implicit in the production of $\b$ and
$\c$ quarks, although the forcing specifically of $\Jpsi$ production
is difficult. In this section are given the main processes for the
third source, intended for applications at hadron colliders.
The traditional `colour singlet' approach is encapsulated in the
above processes in the range 86 -- 108. Processes 104 and 105 are
the equivalents of 87 and 89 in the limit of $\pT \to 0$; note that
$\g \g \to \Jpsi$ and $\g \g \to \chi_{1 \c}$ are forbidden and thus
absent. As always one should beware of double-counting between 87 and
104, and between 89 and 105, and thus use either the one or the other
depending on the kinematic domain to be studied. The cross sections
depend on wave function values at the origin, see \ttt{PARP(38)} and
\ttt{PARP(39)}. A review of the physics issues involved may be found
in \cite{Glo88} (note, however, that the choice of $Q^2$ scale is
different in {\Py}).
While programmed for the charm system, it would be straightforward
to apply these processes instead to bottom mesons, i.e.\ for the
production of $\Upsilon$.One needs to change
the codes of states produced, which is achieved by
\ttt{KFPR(ISUB,1) = KFPR(ISUB,1) + 110} for the processes {\ISUB}
above, and changing the values of the wave functions at the origin,
\ttt{PARP(38)} and \ttt{PARP(39)}.
It is known that the above sources are not enough to explain the
full $\Jpsi$ rate, and further production mechanisms have been
proposed, extending on the more conventional treatment here
\cite{Can97}. The most common extension is the `colour octet'
production mechanism, in the framework of nonrelativistic QCD (NRQCD)
\cite{Bod95}. In this language, production in part proceeds via
intermediate colour octet states that collapse to singlet states
by the emission of soft (and thus nonperturbative) gluons. In the
current implementation \cite{Wol02}, three new colour octet states
are introduced for each of the $\c\cbar$ and $\b\bbar$ systems, with
spectroscopic notation $\Q\Qbar[^{2S+1}L_J^{(8)}]$, where the $(8)$
is a reminder of the colour octet nature of these states. These new
`particles' are assumed to `decay' exclusively to $\Jpsi + \g$
or $\Upsilon + \g$, respectively. Their masses have been chosen to allow
this, without too much excess phase space, so that the emitted gluon
is always very soft.
Unlike the first set of processes above, the NRQCD processes have been
explicitly duplicated for the $\c\cbar$ and $\b\bbar$ sectors. Further,
several processes already present in the colour singlet framework are
repeated here, only differing by the way wave function and matrix element
normalization factors are defined, so as to provide a coherent framework.
For this reason, obviously the processes above 420 should not be combined
with the lower-number ones, or else one would doublecount.
The rates for these new processes are regulated by 10 new NRQCD matrix
element values, \ttt{PARP(141) - PARP(150)}.
%Currently each of those are
%set to unity by default, which obviously is unphysical and will be
%changed soon.
The switches \ttt{MSTP(145) - MSTP(149)} can be used
to modify the behaviour of the processes, but are mainly intended for
experts.
\subsubsection{Minimum bias}
\label{sss:minbiasclass}
\ttt{MSEL} = 1, 2 \\
{\ISUB} =
\begin{tabular}[t]{rl}
91 & elastic scattering \\
92 & single diffraction ($AB \to XB$) \\
93 & single diffraction ($AB \to AX$) \\
94 & double diffraction \\
95 & low-$\pT$ production \\
\end{tabular}
These processes are briefly discussed in section
\ref{ss:nonpertproc}. They are mainly intended for interactions
between hadrons, although one may also consider $\gamma \p$
and $\gamma\gamma$ interactions in the options where the incoming
photon(s) is (are) assumed resolved.
Uncertainties come from a number of sources, e.g.\ from the
parameterizations of the various cross sections and slope parameters.
In diffractive scattering, the structure of the selected hadronic
system may be regulated with \ttt{MSTP(101)}. No high-$\pT$
jet production in diffractive events is included so far; one would have
to use add-on programs like \tsc{PomPyt} \cite{Bru96} for that.
The subprocess 95, low-$\pT$ events, is somewhat unique in
that no meaningful physical border-line to high-$\pT$ events can be
defined. Even if the QCD $2 \to 2$ high-$\pT$ processes are formally
switched off, some of the generated events will be classified as
belonging to this group, with a $\pT$ spectrum of interactions to
match the `minimum-bias' event sample. The generation of such jets
is performed with the help of the auxiliary subprocess 96, see
section \ref{sss:QCDjetclass}. Only with the option
\ttt{MSTP(82) = 0} will subprocess 95 yield strictly low-$\pT$
events, events which will then probably not be compatible with any
experimental data. A number of options exist for the detailed
structure of low-$\pT$ events, see in particular \ttt{MSTP(81)} and
\ttt{MSTP(82)}. Further details on the model(s) for minimum-bias
events are found in sections \ref{ss:multint}--\ref{ss:newmultint}.
\subsection{Physics with Incoming Photons}
\label{ss:photoanddisclass}
With recent additions, the machinery for photon physics has become
rather extensive \cite{Fri00}. The border between the physics of real
photon interactions and of virtual photon ones is now bridged
by a description that continuously interpolates between the two
extremes, as summarized in section \ref{sss:photoprod}.
Furthermore, the {\galep} option (where \textit{lepton} is
to be replaced by \ttt{e-}, \ttt{e+}, \ttt{mu-}, \ttt{mu+},
\ttt{tau-} or \ttt{tau+} as the case may be) in
a \ttt{PYINIT} call gives access to an internally generated spectrum
of photons of varying virtuality. The \ttt{CKIN(61) - CKIN(78)}
variables can be used to set experimentally motivated $x$ and $Q^2$
limits on the photon fluxes. With this option, and the default
\ttt{MSTP(14) = 30}, one automatically obtains a realistic first
approximation to `all' QCD physics of $\gast\p$ and $\gast\gast$
interactions. The word `all' clearly does not mean that a perfect
description is guaranteed, or that all issues are addressed, but
rather that the intention is to simulate all processes that give
a significant contribution to the total cross section in whatever
$Q^2$ range is being studied: jets, low-$\pT$ events, elastic and
diffractive scattering, etc.
The material to be covered encompasses many options, several of
which have been superseded by further developments but have been
retained for backwards compatibility. Therefore it is here split
into three sections. The first covers the physics of real photons
and the subsequent one that of (very) virtual ones. Thereafter, in
the final section, the threads are combined into a machinery
applicable at all $Q^2$.
\subsubsection{Photoproduction and $\gamma\gamma$ physics}
\label{sss:photoprodclass}
\ttt{MSEL} = 1, 2, 4, 5, 6, 7, 8 \\
{\ISUB} =
\begin{tabular}[t]{rl}
33 & $\q_i \gamma \to \q_i \g$ \\
34 & $\f_i \gamma \to \f_i \gamma$ \\
54 & $\g \gamma \to \q_k \qbar_k$ \\
58 & $\gamma \gamma \to \f_k \fbar_k$ \\
80 & $\q_i \gamma \to \q_k \pi^{\pm}$ \\
84 & $\g \gamma \to \Q_k \Qbar_k$ \\
85 & $\gamma \gamma \to \F_k \Fbar_k$ \\
\end{tabular}
An (almost) real photon has both a point-like component and a
hadron-like one. This means that several classes of processes
may be distinguished, see section \ref{sss:photoprod}.
\begin{Enumerate}
\item The processes listed above are possible when the photon
interacts as a point-like particle, i.e.\ couples directly to
quarks and leptons.
\item When the photon acts like a hadron, i.e.\ is resolved in a
partonic substructure, then high-$\pT$ parton--parton interactions
are possible, as described in sections \ref{sss:QCDjetclass}
and \ref{sss:promptgammaclass}. These interactions may be further
subdivided into VMD and anomalous (GVMD) ones \cite{Sch93,Sch93a}.
\item A hadron-like photon can also produce the equivalent of the
minimum-bias processes of section \ref{sss:minbiasclass}. Again,
these can be subdivided into VMD and GVMD (anomalous) ones.
\end{Enumerate}
For $\gamma\p$ events, we believe that the best description can be
obtained when three separate event classes are combined, one for
direct, one for VMD and one for GVMD/anomalous events, see the
detailed description in \cite{Sch93,Sch93a}.
These correspond to \ttt{MSTP(14)} being 0, 2 and 3, respectively.
The direct component is high-$\pT$ only, while VMD and GVMD
contain both high-$\pT$ and low-$\pT$ events. The option
\ttt{MSTP(14) = 1} combines the VMD and GVMD/anomalous parts of the
photon into one single resolved photon concept, which therefore is
less precise than the full subdivision.
When combining three runs to obtain the totality of $\gamma\p$
interactions, to the best of our knowledge, it is necessary to choose
the $\pT$ cut-offs with some care, so as to represent the expected
total cross section.
\begin{Itemize}
\item The direct processes by themselves only depend on the
\ttt{CKIN(3)} cut-off of the generation. In older program versions
the preferred value was 0.5 GeV \cite{Sch93,Sch93a}. In the more recent
description in \cite{Fri00}, also eikonalization of direct with
anomalous interactions into the GVMD event class is considered.
That is, given a branching $\gamma \to \q\qbar$, direct interactions
are viewed as the low-$\pT$ events and anomalous ones as high-$\pT$
events that have to merge smoothly. Then the \ttt{CKIN(3)} cut-off is
increased to the $\pTmin$ of
multiple interactions processes, see \ttt{PARP(81)} (or \ttt{PARP(82)},
depending on minijet unitarization scheme). See \ttt{MSTP(18)} for
a possibility to switch back to the older behaviour. However, full
backwards compatibility cannot be assured, so the older scenarios
are better simulated by using an older {\Py} version.
\item The VMD processes work as ordinary
hadron--hadron ones, i.e.\ one obtains both low- and high-$\pT$
events by default, with dividing line set by $\pTmin$ above.
\item Also the GVMD processes work like the VMD ones. Again this is
a change from previous versions, where the anomalous processes only
contained high-$\pT$ physics and the low-$\pT$ part was covered in the
direct event class. See \ttt{MSTP(15) = 5} for a possibility to switch
back to the older behaviour, with comments as above for the direct
class. A GVMD state is book-kept as a diffractive state in the event
listing, even when it scatters `elastically', since the subsequent
hadronization descriptions are very similar.
\end{Itemize}
The processes in points 1 and 2 can be simulated with a photon beam,
i.e.\ when \ttt{'gamma'} appears as argument in the \ttt{PYINIT} call.
It is then necessary to use option \ttt{MSTP(14)} to switch between
a point-like and a resolved photon --- it is not possible to simulate
the two sets of processes in a single run. This would be the normal
mode of operation for beamstrahlung photons, which have $Q^2 = 0$
but with a nontrivial energy spectrum that would be provided by some
external routine.
For bremsstrahlung photons, the $x$ and $Q^2$ spectrum can be simulated
internally, with the {\galep} argument in the \ttt{PYINIT}
call. This is the recommended procedure, wherein direct and resolved
processes can be mixed. An older --- now not recommended --- alternative
is to use a parton-inside-electron structure function concept, obtainable
with a simple \ttt{'e-'} (or other lepton) argument in \ttt{PYINIT}.
To access these quark and gluon distributions inside the photon (itself
inside the electron), \ttt{MSTP(12) = 1} must then be used. Also the
default value \ttt{MSTP(11) = 1} is required for the preceding step, that
of finding photons inside the electron. Also here the direct and resolved
processes may be generated together. However, this option only works
for high-$\pT$ physics. It is not possible to have also the low-$\pT$
physics (including multiple interactions in high-$\pT$ events) for an
electron beam. Kindly note that subprocess 34 contains both the scattering
of an electron off a photon and the scattering of a quark (inside a photon
inside an electron) off a photon; the former can be switched off with the
help of the \ttt{KFIN} array.
If you are only concerned with standard QCD physics, the option
\ttt{MSTP(14) = 10} or the default \ttt{MSTP(14) = 30} gives an automatic
mixture of the VMD, direct and GVMD/ano\-ma\-lous event classes.
The mixture is properly given according to
the relative cross sections. Whenever possible, this option is therefore
preferable in terms of user-friendliness. However, it can only work
because of a completely new layer of administration, not found anywhere
else in {\Py}. For instance, a subprocess like $\q\g \to \q\g$ is
allowed in several of the classes, but appears with different sets of
parton distributions and different $\pT$ cut-offs in each of these,
so that it is necessary to switch gears between each event in the
generation. It is therefore not possible to avoid a number of
restrictions on what you can do in this case:
\begin{Itemize}
\item The \ttt{MSTP(14) = 10} and \ttt{= 30} options can only be used for
incoming photon beams, with or without convolution with the
bremsstrahlung spectrum, i.e.\ when \ttt{'gamma'} or {\galep}
is the argument in the \ttt{PYINIT} call.
\item The machinery has only been set up to generate standard
QCD physics, specifically either `minimum-bias' one or high-$\pT$ jets.
There is thus no automatic mixing of processes only for heavy-flavour
production, say, or of some exotic particle.
For minimum bias, you are not allowed to use the \ttt{CKIN} variables
at all. This is not a major limitation, since it is in the spirit of
minimum-bias physics not to impose any constraints on allowed jet
production. (If you still do, these cuts will be ineffective for the
VMD processes but take effect for the other ones, giving
inconsistencies.)
The minimum-bias physics option is obtained by default; by switching
from \ttt{MSEL = 1} to \ttt{MSEL = 2} also the elastic and diffractive
components of the VMD and GVMD parts are included. High-$\pT$ jet
production is obtained by setting the \ttt{CKIN(3)} cut-off larger than
the $\pTmin(W^2)$ of the multiple interactions scenario. For
lower input \ttt{CKIN(3)} values the program will automatically switch
back to minimum-bias physics.
\item Multiple interactions become possible in both the VMD and GVMD
sector, with the average number of interactions given by the
ratio of the jet to the total cross section. Currently only
the simpler scenario \ttt{MSTP(82) = 1} in the old model is implemented,
however, i.e.\ the more sophisticated variable-impact-parameter ones
need further physics studies and model development.
\item Some variables are internally recalculated and reset, notably
\ttt{CKIN(3)}. This is because it must have values that depend on the
component studied. It can therefore not be modified without changing
\ttt{PYINPR} and recompiling the program, which obviously is a major
exercise.
\item Pileup events are not at all allowed.
\end{Itemize}
Also, a warning about the usage of \tsc{Pdflib}/\tsc{LHAPDF} for
photons. So long as \ttt{MSTP(14) = 1}, i.e.\ the photon is not split up,
\tsc{Pdflib} is accessed by \ttt{MSTP(56) = 2} and \ttt{MSTP(55)} as the
parton distribution set. However, when the VMD and anomalous pieces are
split, the VMD part is based on a rescaling of pion distributions by VMD
factors (except for the SaS sets, that already come with a separate VMD
piece). Therefore, to access \tsc{Pdflib} for \ttt{MSTP(14) = 10}, it is
not correct to set \ttt{MSTP(56) = 2} and a photon distribution in
\ttt{MSTP(55)}. Instead, one should put \ttt{MSTP(56) = 2},
\ttt{MSTP(54) = 2} and a pion distribution code in \ttt{MSTP(53)},
while \ttt{MSTP(55)} has no function. The anomalous part is still based on
the SaS parameterization, with \ttt{PARP(15)} as main free parameter.
Currently, hadrons are not defined with any photonic content. None
of the processes are therefore relevant in hadron--hadron collisions.
In $\e\p$ collisions, the electron can emit an almost real photon,
which may interact directly or be resolved. In $\ee$ collisions,
one may have direct, singly-resolved or doubly-resolved processes.
The $\gamma\gamma$ equivalent to the $\gamma\p$ description involves
six different event classes, see section \ref{sss:photoprod}.
These classes can be obtained by setting \ttt{MSTP(14)} to 0, 2, 3,
5, 6 and 7, respectively. If one combines the VMD and anomalous
parts of the parton distributions of the photon, in a more coarse
description, it is enough to use the \ttt{MSTP(14)} options 0, 1 and 4.
The cut-off procedures follows from the ones used for the $\gamma\p$
ones above.
As with $\gamma\p$ events, the options \ttt{MSTP(14) = 10} or
\ttt{MSTP(14) = 30} give a mixture of the six possible $\gamma\gamma$
event classes. The same complications and restrictions exist here as
already listed above.
Process 54 generates a mixture of quark flavours; allowed flavours
are set by the gluon \ttt{MDME values}. Process 58 can generate both
quark and lepton pairs, according to the \ttt{MDME} values of the
photon. Processes 84 and 85 are variants of these matrix elements,
with fermion masses included in the matrix elements, but where only
one flavour can be generated at a time. This flavour is selected as
described for processes 81 and 82 in section \ref{sss:heavflavclass},
with the exception that for process 85 the `heaviest' flavour allowed
for photon splitting takes to place of the heaviest flavour allowed
for gluon splitting. Since lepton {\KF} codes come after quark ones,
they are counted as being `heavier', and thus take precedence if
they have been allowed.
Process 80 is a higher twist one. The theory for such processes
is rather shaky, so results should not be taken too literally.
The messy formulae given in \cite{Bag82} have not been programmed
in full, instead the pion form factor has been parameterized as
$Q^2 F_{\pi}(Q^2) \approx 0.55 / \ln Q^2$, with $Q$ in GeV.
\subsubsection{Deeply Inelastic Scattering and $\gamma^*\gamma^*$ physics}
\label{sss:DISclass}
\ttt{MSEL} = 1, 2, 35, 36, 37, 38 \\
{\ISUB} =
\begin{tabular}[t]{rl}
10 & $\f_i \f_j \to \f_k \f_l$ \\
83 & $\q_i \f_j \to \Q_k \f_l$ \\
99 & $\gast \q \to \q$ \\
131 & $\f_i \gast_{\mrm{T}} \to \f_i \g$ \\
132 & $\f_i \gast_{\mrm{L}} \to \f_i \g$ \\
133 & $\f_i \gast_{\mrm{T}} \to \f_i \gamma$ \\
134 & $\f_i \gast_{\mrm{L}} \to \f_i \gamma$ \\
135 & $\g \gast_{\mrm{T}} \to \f_i \fbar_i$ \\
136 & $\g \gast_{\mrm{L}} \to \f_i \fbar_i$ \\
137 & $\gast_{\mrm{T}} \gast_{\mrm{T}} \to \f_i \fbar_i$ \\
138 & $\gast_{\mrm{T}} \gast_{\mrm{L}} \to \f_i \fbar_i$ \\
139 & $\gast_{\mrm{L}} \gast_{\mrm{T}} \to \f_i \fbar_i$ \\
140 & $\gast_{\mrm{L}} \gast_{\mrm{L}} \to \f_i \fbar_i$ \\
\end{tabular}
Among the processes in this section, 10 and 83 are intended to
stand on their own, while the rest are part of the newer machinery
for $\gast\p$ and $\gast\gast$ physics. We therefore separate
the description in this section into these two main parts.
The Deeply Inelastic Scattering (DIS) processes, i.e.\ $t$-channel
electroweak gauge boson exchange, are traditionally associated
with interactions between a lepton or neutrino and a hadron, but
processes 10 and 83 can equally well be applied for $\q\q$ scattering
in hadron colliders (with a cross section much smaller than
corresponding QCD processes, however). If applied to incoming $\ee$
beams, process 10 corresponds to Bhabha scattering.
For process 10 both $\gamma$, $\Z^0$ and $\W^{\pm}$ exchange
contribute, including interference between $\gamma$ and $\Z^0$.
The switch \ttt{MSTP(21)} may be used to restrict to only some
of these, e.g.\ neutral or charged current only.
The option \ttt{MSTP(14) = 10} (see previous section) has now been
extended so that it also works for DIS of
an electron off a (real) photon, i.e.\ process 10. What is obtained
is a mixture of the photon acting as a vector meson and it acting
as an anomalous state. This should therefore be the sum of what can
be obtained with \ttt{MSTP(14) = 2} and \ttt{= 3}. It is distinct from
\ttt{MSTP(14) = 1} in that different sets are used for the parton
distributions --- in \ttt{MSTP(14) = 1} all the contributions to the
photon distributions are lumped together, while they are split in
VMD and anomalous parts for \ttt{MSTP(14) = 10}. Also the beam-remnant
treatment is different, with a simple Gaussian distribution (at least
by default) for \ttt{MSTP(14) = 1} and the VMD part of
\ttt{MSTP(14) = 10}, but a powerlike distribution
$\d k_{\perp}^2 / k_{\perp}^2$ between \ttt{PARP(15)} and $Q$ for
the anomalous part of \ttt{MSTP(14) = 10}.
To access this option for $\e$ and $\gamma$ as incoming beams, it is
only necessary to set \ttt{MSTP(14) = 10} and keep \ttt{MSEL} at its
default value. Unlike the corresponding option for $\gamma\p$ and
$\gamma\gamma$, no cuts are overwritten, i.e.\ it is still your
responsibility to set these appropriately.
Cuts especially appropriate for DIS usage include either
\ttt{CKIN(21) - CKIN(22)} or \ttt{CKIN(23) - CKIN(24)} for the $x$
range (former or latter depending on which side is the incoming
real photon), \ttt{CKIN(35) - CKIN(36)} for
the $Q^2$ range, and \ttt{CKIN(39) - CKIN(40)} for the $W^2$ range.
In principle, the DIS $x$ variable of an event corresponds to the
$x$ value stored in \ttt{PARI(33)} or \ttt{PARI(34)}, depending
on which side the incoming hadron is on, while the DIS
$Q^2 = -\hat{t} = $\ttt{-PARI(15)}. However, just like initial- and
final-state radiation can shift jet momenta, they can modify
the momentum of the scattered lepton. Therefore the DIS $x$ and
$Q^2$ variables are not automatically conserved. An option, on by
default, exists in \ttt{MSTP(23)}, where the event can be `modified
back' so as to conserve $x$ and $Q^2$, but this option is rather
primitive and should not be taken too literally.
Process 83 is the equivalent of process 10 for $\W^{\pm}$ exchange
only, but with the heavy-quark mass included in the matrix element.
In hadron colliders it is mainly of interest for the production of
very heavy flavours, where the possibility of producing just one
heavy quark is kinematically favoured over pair production. The
selection of the heavy flavour is already discussed in section
\ref{sss:heavflavclass}.
Turning to the other processes, part of the $\gast\p$ and $\gast\gast$
process-mixing machineries, 99 has close similarities with the
above discussed 10 one. Whereas 10 would simulate the full process
$\e \q \to \e \q$, 99 assumes a separate machinery for the flux
of virtual photons, $\e \to \e \gast$ and only covers the second
half of the process, $\gast \q \to \q$. One limitation of this
factorization is that only virtual photons are considered in
process 99, not contributions from the $\Z^0$ neutral current
or the $\W^{\pm}$ charged current.
Note that 99 has no correspondence in the real-photon case, but has
to vanish in this limit by gauge invariance, or indeed by simple
kinematics considerations. This, plus the desire to avoid double-counting
with real-photon physics processes, is why the cross section for
this process is explicitly made to vanish for photon virtuality
$Q^2 \to 0$, eq.~(\ref{eq:sigDIS}), also when parton distributions
have not been constructed to fulfil this, see \ttt{MSTP(19)}.
(No such safety measures are present in 10, again illustrating how
the two are intended mainly to be used at large or at small $Q^2$,
respectively.)
For a virtual photon, processes 131--136 may be viewed as first-order
corrections to 99. The three with a transversely polarized photon,
131, 133 and 135, smoothly reduce to the real-photon direct
(single-resolved for $\gamma\gamma$) processes 33, 34 and 54.
The other three, corresponding to the exchange of a longitudinal
photon, vanish like $Q^2$ for $Q^2 \to 0$. The double-counting issue
with process 99 is solved by requiring the latter process not to
contain any shower branchings with a $\pT$ above the lower $\pT$
cut-off of processes 131-136. The cross section is then to be reduced
accordingly, see eq.~(\ref{eq:LODIS}) and the discussion there,
and again \ttt{MSTP(19)}.
We thus see that process 99 by default is a low-$\pT$ process in
about the same sense as process 95, giving `what is left' of the total
cross section when jet events have been removed. Therefore, it will be
switched off in event class mixes such as \ttt{MSTP(14) = 30} if
\ttt{CKIN(3)} is above $\pTmin(W^2)$ and \ttt{MSEL} is not 2. There
is a difference, however, in that process 99 events still are allowed
to contain shower evolution (although currently only the final-state
kind has been implemented), since the border to the other processes
is at $\pT = Q$ for large $Q$ and thus need not be so small. The $\pT$
scale of the `hard process', stored e.g.\ in \ttt{PARI(17)} always
remains 0, however. (Other \ttt{PARI} variables defined for normal
$2 \to 2$ and $2 \to 1$ processes are not set at all, and may well
contain irrelevant junk left over from previous events.)
Processes 137--140, finally, are extensions of process 58 from
the real-photon limit to the virtual-photon case, and correspond to
the direct process of $\gamma^*\gamma^*$ physics. The four cases
correspond to either of the two photons being either transversely or
longitudinally polarized. As above, the cross section of a longitudinal
photon vanishes when its virtuality approaches 0.
\subsubsection{Photon physics at all virtualities}
\label{sss:photonallQclass}
{\ISUB} =
\begin{tabular}[t]{ll}
direct$\times$direct: & 137, 138, 139, 140 \\
direct$\times$resolved: & 131, 132, 135, 136 \\
DIS$\times$resolved: & 99 \\
resolved$\times$resolved, high-$\pT$: & 11, 12, 13, 28, 53, 68 \\
resolved$\times$resolved, low-$\pT$: & 91, 92, 93, 94, 95 \\
\end{tabular}\\
where `resolved' is a hadron or a VMD or GVMD photon.
At intermediate photon virtualities, processes described in both of
the sections above are allowed, and have to be mixed appropriately.
The sets are of about equal importance at around
$Q^2 \sim m_{\rho}^2 \sim 1$~GeV$^2$, but the transition is gradual
over a larger $Q^2$ range. The ansatz for this mixing is given by
eq.~(\ref{eq:gammapallQ}) for $\gast\p$ events and
eq.~(\ref{eq:gagaallQ}) for $\gast\gast$ ones. In short, for direct
and DIS processes the photon virtuality explicitly enters in the
matrix element expressions, and thus is easily taken into account.
For resolved photons, perturbation theory does not provide a unique
answer, so instead cross sections are suppressed by dipole factors,
$(m^2/(m^2 + Q^2))^2$, where $m = m_V$ for a VMD state and $m = 2 \kT$
for a GVMD state characterized by a $\kT$ scale of the
$\gast \to \q\qbar$ branching. These factors appear explicitly for
total, elastic and diffractive cross sections, and are also implicitly
used e.g.\ in deriving the SaS parton distributions for virtual photons.
Finally, some double-counting need to be removed, between direct and
DIS processes as mentioned in the previous section, and between
resolved and DIS at large $x$.
Since the mixing is not trivial, it is recommended to use the default
\ttt{MSTP(14) = 30} to obtain it in one go and hopefully consistently,
rather than building it up by combining separate runs. The main issues
still under your control include, among others
\begin{Itemize}
\item The \ttt{CKIN(61) - CKIN(78)} should be used to set the range of
$x$ and $Q^2$ values emitted from the lepton beams. That way one may
decide between almost real or very virtual photons, say. Also some other
quantities, like $W^2$, can be constrained to desirable ranges.
\item Whether or not minimum-bias events are simulated depends on the
\ttt{CKIN(3)} value, just like in hadron physics. The only difference is
that the initialization energy scale $W_{\mrm{init}}$ is selected in the
allowed $W$ range rather than to be the full c.m.\ energy.\\
For a high \ttt{CKIN(3)}, \ttt{CKIN(3)} $> \pTmin(W_{\mrm{init}}^2)$,
only jet production is included. Then further \ttt{CKIN} values can be
set to constrain e.g.\ the rapidity of the jets produced.\\
For a low \ttt{CKIN(3)}, \ttt{CKIN(3)} $< \pTmin(W_{\mrm{init}}^2)$,
like the default value \ttt{CKIN(3) = 0}, low-$\pT$ physics is switched
on together with jet production, with the latter properly eikonalized to
be lower than the total one. The ordinary \ttt{CKIN} cuts, not related to
the photon flux, cannot be used here.\\
For a low \ttt{CKIN(3)}, when \ttt{MSEL = 2} instead of the default
\ttt{= 1}, also elastic and diffractive events are simulated.
\item The impact of resolved longitudinal photons is not unambiguous,
e.g.\ only recently the first parameterization of parton distributions
appeared \cite{Chy00}. Different simple alternatives can be probed by
changing \ttt{MSTP(17)} and associated parameters.
\item The choice of scales to use in parton distributions for jet rates
is always ambiguous, but depends on even more scales for virtual photons
than in hadronic collisions. \ttt{MSTP(32)} allows a choice between
several alternatives.
\item The matching of $\pT$ generation by shower evolution to that by
primordial $\kT$ is a general problem, for photons with an additional
potential source in the $\gast \to \q\qbar$ vertex. \ttt{MSTP(66)}
offer some alternatives.
\item \ttt{PARP(15)} is the $k_0$ parameter separating VMD from GVMD.
\item \ttt{PARP(18)} is the $k_{\rho}$ parameter in GVMD total cross
sections.
\item \ttt{MSTP(16)} selects the momentum variable for an
$\e \to \e \gast$ branching.
\item \ttt{MSTP(18)} regulates the choice of $\pTmin$ for direct
processes.
\item \ttt{MSTP(19)} regulates the choice of partonic cross section in
process 99, $\gast\q \to \q$.
\item \ttt{MSTP(20)} regulates the suppression of the resolved cross
section at large $x$.
\end{Itemize}
The above list is not complete, but gives some impression what can
be done.
\subsection{Electroweak Gauge Bosons}
This section covers the production and/or exchange of $\gamma$,
$\Z^0$ and $\W^{\pm}$ gauge bosons, singly and in pairs. The topic
of longitudinal gauge-boson scattering at high energies is deferred
to the Higgs section, since the presence or absence of a Higgs boson here
makes a big difference.
\subsubsection{Prompt photon production}
\label{sss:promptgammaclass}
\ttt{MSEL} = 10 \\
{\ISUB} =
\begin{tabular}[t]{rl}
14 & $\q_i \qbar_i \to \g \gamma$ \\
18 & $\f_i \fbar_i \to \gamma \gamma$ \\
29 & $\q_i \g \to \q_i \gamma$ \\
114 & $\g \g \to \gamma \gamma$ \\
115 & $\g \g \to \g \gamma$ \\
\end{tabular}
In hadron colliders, processes {\ISUB} = 14 and 29 give the main source
of single-$\gamma$ production, with {\ISUB} = 115 giving an additional
contribution which, in some kinematics regions, may become important.
For $\gamma$-pair production, the process {\ISUB} = 18 is often
overshadowed in importance by {\ISUB} = 114.
Another source of photons is bremsstrahlung off incoming or outgoing
quarks. This has to be treated on an equal footing with QCD parton
showering. For time-like parton-shower evolution, i.e.\ in the
final-state showering and in the side branches of the initial-state
showering, photon emission may be switched on or off with
\ttt{MSTJ(41)}. Photon radiation off the space-like incoming
quark or lepton legs is similarly regulated by \ttt{MSTP(61)}.
{\bf Warning:} the cross sections for the box graphs 114 and 115 become
very complicated, numerically unstable and slow when the
full quark mass dependence is included. For quark masses much
below the $\hat{s}$ scale, the simplified massless expressions are
therefore used --- a fairly accurate approximation. However, there
is another set of subtle numerical cancellations between different
terms in the massive matrix elements in the region of small-angle
scattering. The associated problems have not been sorted out yet.
There are therefore two possible solutions. One is to use the
massless formulae throughout. The program then becomes faster and
numerically stable, but does not give, for example, the characteristic
dip (due to destructive interference) at top threshold. This is the
current default procedure, with five flavours assumed, but this
number can be changed in \ttt{MSTP(38)}. The other possibility is
to impose cuts on the scattering angle of the hard process, see
\ttt{CKIN(27)} and \ttt{CKIN(28)}, since the numerically unstable
regions are when $|\cos\hat{\theta}|$ is close to unity. It is then
also necessary to change \ttt{MSTP(38)} to 0.
\subsubsection{Single $\W/\Z$ production}
\label{sss:WZclass}
\ttt{MSEL} = 11, 12, 13, 14, 15, (21) \\
{\ISUB} =
\begin{tabular}[t]{rl}
1 & $\f_i \fbar_i \to \gammaZ$ \\
2 & $\f_i \fbar_j \to \W^+$ \\
15 & $\f_i \fbar_i \to \g (\gammaZ)$ \\
16 & $\f_i \fbar_j \to \g \W^+$ \\
19 & $\f_i \fbar_i \to \gamma (\gammaZ)$ \\
20 & $\f_i \fbar_j \to \gamma \W^+$ \\
30 & $\f_i \g \to \f_i (\gammaZ)$ \\
31 & $\f_i \g \to \f_k \W^+$ \\
35 & $\f_i \gamma \to \f_i (\gammaZ)$ \\
36 & $\f_i \gamma \to \f_k \W^+$ \\
(141) & $\f_i \fbar_i \to \gamma/\Z^0/\Z'^0$ \\
(142) & $\f_i \fbar_j \to \W'^+$ \\
\end{tabular}
This group consists of $2 \to 1$ processes, i.e.\ production of a
single resonance, and $2 \to 2$ processes, where the resonance
is recoiling against a jet or a photon. The processes 141 and 142,
which also are listed here, are described further elsewhere.
With initial-state showers turned on, the $2 \to 1$ processes also
generate additional jets; in order to avoid double-counting, the
corresponding $2 \to 2$ processes should therefore not be turned
on simultaneously. The basic rule is to use the $2 \to 1$ processes
for inclusive generation of $\W/\Z$, i.e.\ where the bulk of the
events studied have $\pT \ll m_{\W/\Z}$. With the introduction of
explicit matrix-element-inspired corrections to the parton shower
\cite{Miu99}, also the high-$\pT$ tail is well described in this
approach, thus offering an overall good description of the full $\pT$
spectrum of gauge bosons \cite{Bal01}.
If one is interested in the high-$\pT$ tail only, however, the
generation efficiency will be low. It is here better to start from
the $2 \to 2$ matrix elements and add showers to these. However, the
$2 \to 2$ matrix elements are divergent for $\pT \to 0$, and should
not be used down to the low-$\pT$ region, or one may get unphysical
cross sections. As soon as the generated $2 \to 2$ cross section
corresponds to a non-negligible fraction of the total $2 \to 1$ one,
say 10\%--20\%, Sudakov effects are likely to be affecting the
shape of the $\pT$ spectrum to a corresponding extent, and results
should not be trusted.
The problems of double-counting and Sudakov effects apply not only
to $\W/\Z$ production in hadron colliders, but also to a process
like $\ee \to \Z^0 \gamma$, which clearly is part of the
initial-state radiation corrections to $\ee \to \Z^0$ obtained for
\ttt{MSTP(11) = 1}. As is the case for $\Z$ production in association
with jets, the $2 \to 2$ process should therefore only be used for the
high-$\pT$ region.
The $\Z^0$ of subprocess 1 includes the full interference structure
$\gammaZ$; via \ttt{MSTP(43)} you can select to produce only
$\gamma^*$, only $\Z^0$, or the full $\gammaZ$. The same holds true
for the $\Z'^0$ of subprocess 141; via \ttt{MSTP(44)} any combination
of $\gamma^*$, $\Z^0$ and $\Z'^0$ can be selected. Thus, subprocess
141 with \ttt{MSTP(44) = 4} is essentially equivalent to subprocess
1 with \ttt{MSTP(43) = 3}; however, process 141 also includes the
possibility of a decay into Higgs bosons. Also processes
15, 19, 30 and 35 contain the full mixture of $\gammaZ$, with
\ttt{MSTP(43)} available to change this. Note that the
$\gammaZ$ decay products can have an invariant mass as small
as the program cutoff. This can be changed using \ttt{CKIN}.
Note that process 1, with only
$\q\qbar \to \gamma^* \to \ell^+ \ell^-$ allowed,
and studied in the region well below the $\Z^0$ mass, is what is
conventionally called Drell--Yan. This latter process therefore does
not appear under a separate heading, but can be obtained by a
suitable setting of switches and parameters.
A process like $\f_i \fbar_j \to \gamma \W^+$ requires some comment.
When the $\W$ boson decays, photons can be radiated off the decay
products. The full interference between photon radiation off
the incoming fermions, the intermediate $\W$ boson, and the decay
products is not included in the $\Py$ treatment. If such
effects are important, a full matrix element calculation is
preferred. Some caution must therefore be
exercised; see also section \ref{sss:WZpairclass} for related
comments.
For the $2 \to 1$ processes, the Breit--Wigner includes an
$\hat{s}$-dependent width, which should provide an improved
description of line shapes. In fact, from a line-shape point of view,
process 1 should provide a more accurate simulation of $\ee$
annihilation events than the dedicated $\ee$ generation scheme of
\ttt{PYEEVT} (see section \ref{ss:eematrix}). Another difference is
that \ttt{PYEEVT} only allows the generation of $\gammaZ \to \q\qbar$,
while process 1 additionally contains $\gammaZ \to \ell^+ \ell^-$ and
$\nu \br{\nu}$. The parton-shower and fragmentation
descriptions are the same, but the process 1 implementation only
contains a partial interface to the first- and second-order
matrix-element options available in \ttt{PYEEVT}, see \ttt{MSTP(48)}.
All processes in this group have been included with the
correct angular distribution in the subsequent $\W/\Z \to \f \fbar$
decays. In process 1 also fermion mass effects
have been included in the angular distributions, while this is not the
case for the other ones. Normally mass effects are not large anyway.
As noted earlier, some approximations can be used to simulate higher-order
processes.
The process $\ee \to \e^+ \e^- \Z^0$ can be simulated in two
different ways. One is to make use of the $\e$ `sea' distribution
inside $\e$, i.e.\ have splittings $\e \to \gamma \to \e$.
This can be obtained, together with ordinary $\Z^0$ production, by
using subprocess 1, with \ttt{MSTP(11) = 1} and \ttt{MSTP(12) = 1}. Then
the contribution of the type above is 5.0 pb for a 500 GeV $\ee$
collider, compared with the correct 6.2 pb \cite{Hag91}. Alternatively
one may use process 35, with \ttt{MSTP(11) = 1} and \ttt{MSTP(12) = 0},
relying on the splitting $\e \to \gamma$.
This process has a singularity in the forward direction, regularized by
the electron mass and also sensitive to the virtuality of the photon.
It is therefore among the few where the incoming masses have been
included in the matrix element expression. Nevertheless, it may be
advisable to set small lower cut-offs, e.g.\
\ttt{CKIN(3) = CKIN(5) = 0.01},
if one should experience problems (e.g.\ at higher energies).
Process 36, $\f \gamma \to \f' \W^{\pm}$ may have corresponding
problems; except that in $\ee$ the forward scattering amplitude for
$\e \gamma \to \nu \W$ is killed (radiation zero), which means
that the differential cross section is vanishing for $\pT \to 0$.
It is therefore feasible to use the default \ttt{CKIN(3)} and
\ttt{CKIN(5)} values in $\ee$, and one also comes closer to the
correct cross section.
The process $\g \g \to \Z^0 \b \bbar$, formerly available as process
131, has been removed from the current version, since the implementation
turned out to be slow and unstable. However, process 1 with incoming
flavours set to be $\b \bbar$ (by
\ttt{KFIN(1,5) = KFIN(1,-5) = KFIN(2,5) = KFIN(2,-5) = 1} and everything
else \ttt{= 0}) provides an alternative description, where the
additional $\b \bbar$ are generated by $\g \to \b \bbar$ branchings
in the initial-state showers. (Away from the low-$\pT$ region,
process 30 with \ttt{KFIN} values as above except that also incoming
gluons are allowed, offers yet another description. Here it is in terms
of $\g\b \to \Z^0 \b$, with only one further $\g \to \b \bbar$ branching
constructed by the shower.) At first glance, the shower approach would
seem less reliable than the full $2 \to 3$ matrix element. The relative
lightness of the $\b$ quark will generate large logs of the type
$\ln(m_{\Z}^2/m_{\b}^2)$, however, that ought to be resummed \cite{Car00}.
This is implicit in the parton-density approach of incoming $\b$ quarks
but absent from the lowest-order $\g \g \to \Z^0 \b \bbar$ matrix elements.
Therefore actually the shower approach may be the more accurate of the
two in the region of intermediate transverse momenta.
\subsubsection{$\W/\Z$ pair production}
\label{sss:WZpairclass}
\ttt{MSEL} = 15 \\
{\ISUB} =
\begin{tabular}[t]{rl}
22 & $\f_i \fbar_i \to (\gammaZ) (\gammaZ)$ \\
23 & $\f_i \fbar_j \to \Z^0 \W^+$ \\
25 & $\f_i \fbar_i \to \W^+ \W^-$ \\
69 & $\gamma \gamma \to \W^+ \W^-$ \\
70 & $\gamma \W^+ \to \Z^0 \W^+$ \\
\end{tabular}
In this section we mainly consider the production of $\W/\Z$ pairs
by fermion--antifermion annihilation, but also include two processes
which involve $\gamma/\W$ beams. Scatterings between gauge-boson
pairs, i.e.\ processes like $\W^+ \W^- \to \Z^0 \Z^0$, depend so
crucially on the assumed Higgs scenario that they are considered
separately in section \ref{sss:heavySMHclass}.
The cross sections used for the above processes are those derived
in the narrow-width limit, but have been extended to include
Breit--Wigner shapes with mass-dependent widths for the final-state
particles. In process 25, the contribution from $\Z^0$ exchange
to the cross section is now evaluated with the fixed nominal $\Z^0$
mass and width in the propagator. If instead the actual mass and the
%running width were to be used, it gives a diverging cross section at
%referee: misprint
running width were to be used, it would give a diverging cross section at
large energies, by imperfect gauge cancellation.
However, one should realize that other graphs, not included here, can
contribute in regions away from the $\W/\Z$ mass. This problem is
especially important if several flavours coincide in the four-fermion
final state. Consider, as an example,
$\ee \to \mu^+ \mu^- \nu_{\mu} \br{\nu}_{\mu}$.
Not only would such a final state receive contributions from
intermediate $\Z^0\Z^0$ and $\W^+\W^-$ states, but also
from processes $\ee \to \Z^0 \to \mu^+ \mu^-$, followed
either by $\mu^+ \to \mu^+ \Z^0 \to \mu^+ \nu_{\mu} \br{\nu}_{\mu}$,
or by
$\mu^+ \to \br{\nu}_{\mu} \W^+ \to \br{\nu}_{\mu} \mu^+ \nu_{\mu}$.
In addition, all possible interferences should be considered.
Since this is not done, the processes have to be used with some
sound judgement. Very often, one may wish to constrain a
lepton pair mass to be close to $m_{\Z}$, in which case a number
of the possible `other' processes are negligible.
For the $\W$ pair production graph, one experimental objective is to do
precision measurements of the cross section near threshold. Then also
other effects enter. One such is Coulomb corrections, induced by
photon exchange between the two $\W$'s and their decay products.
The gauge invariance issues induced by the finite $\W$ lifetime are not
yet fully resolved, and therefore somewhat different approximate formulae
may be derived \cite{Kho96}. The options in \ttt{MSTP(40)} provide a
reasonable range of uncertainty.
Of the above processes, the first contains the full
$\f_i \fbar_i \to (\gammaZ)(\gammaZ)$ structure, obtained by a
straightforward generalization of the formulae in ref. \cite{Gun86}
(done by one of the {\Py} authors). Of course, the possibility of
there being significant contributions from graphs that are not
included is increased, in particular
if one $\gamma^*$ is very light and therefore could be a
bremsstrahlung-type photon. It is possible to use \ttt{MSTP(43)} to
recover the pure $\Z^0$ case, i.e.\ $\f_i \fbar_i \to \Z^0 \Z^0$
exclusively. In processes 23 and 70, only the pure $\Z^0$ contribution
is included.
Full angular correlations are included for the first three processes,
i.e.\ the full $2 \to 2 \to 4$ matrix elements are included in the
resonance decays, including the appropriate $\gammaZ$ interference
in process 22. In the latter two processes, 69 and 70, no spin
information is currently preserved, i.e.\ the $\W/\Z$ bosons are
allowed to decay isotropically.
We remind you that the mass ranges of the two resonances may be set
with the \ttt{CKIN(41) - CKIN(44)} parameters; this is particularly
convenient, for instance, to pick one resonance almost on the mass
shell and the other not.
\subsection{Higgs Production}
\label{ss:Hclass}
A fair fraction of all the processes in {\Py} deal with Higgs
production in one form or another. This multiplication is caused by
the need to consider production by several different mechanisms,
depending on Higgs mass and machine type. Further, the program
contains a full two-Higgs-multiplet scenario, as predicted for example
in the Minimal Supersymmetric extension of the Standard Model
(MSSM). Therefore the continued discussion is, somewhat arbitrarily,
subdivided into a few different scenarios. Doubly-charged Higgs
particles appear in left--right symmetric models, and are covered in
section \ref{sss:LRDCHclass}.
\subsubsection{Light Standard Model Higgs}
\label{sss:lightSMHclass}
\ttt{MSEL} = 16, 17, 18 \\
{\ISUB} =
\begin{tabular}[t]{rl}
3 & $\f_i \fbar_i \to \hrm^0$ \\
24 & $\f_i \fbar_i \to \Z^0 \hrm^0$ \\
26 & $\f_i \fbar_j \to \W^+ \hrm^0$ \\
32 & $\f_i \g \to \f_i \hrm^0$ \\
102 & $\g \g \to \hrm^0$ \\
103 & $\gamma \gamma \to \hrm^0$ \\
110 & $\f_i \fbar_i \to \gamma \hrm^0$ \\
111 & $\f_i \fbar_i \to \g \hrm^0$ \\
112 & $\f_i \g \to \f_i \hrm^0$ \\
113 & $\g \g \to \g \hrm^0$ \\
121 & $\g \g \to \Q_k \Qbar_k \hrm^0$ \\
122 & $\q_i \qbar_i \to \Q_k \Qbar_k \hrm^0$ \\
123 & $\f_i \f_j \to \f_i \f_j \hrm^0$ ($\Z^0 \Z^0$ fusion) \\
124 & $\f_i \f_j \to \f_k \f_l \hrm^0$ ($\W^+ \W^-$ fusion) \\
\end{tabular}
In this section we discuss the production of a reasonably light
Standard Model Higgs, below 700 GeV, say, so that the narrow
width approximation can be used with some confidence. Below 400 GeV
there would certainly be no trouble, while above that the narrow
width approximation is gradually starting to break down.
In a hadron collider, the main production processes are 102, 123
and 124, i.e.\ $\g\g$, $\Z^0 \Z^0$ and $\W^+ \W^-$ fusion. In the
latter two processes, it is also necessary to take into account
the emission of the space-like $\W/\Z$ bosons off quarks, which
in total gives the $2 \to 3$ processes above.
Other processes with lower cross sections may be of interest
because they provide signals with less background. For instance,
processes 24 and 26 give
associated production of a $\Z$ or a $\W$ together with the $\hrm^0$.
There is also the processes 3 (see below), 121 and 122, which involve
production of heavy flavours.
Process 3 contains contributions from all flavours, but is
completely dominated by the subprocess $\t \tbar \to \hrm^0$,
i.e.\ by the contribution from the top sea distributions.
This assumes that parton densities for top quarks are provided,
which is no longer the case in current parameterizations of PDF's.
This process is by now known to overestimate the cross section
for Higgs production as compared with a more careful calculation
based on the subprocess $\g \g \to \t \tbar \hrm^0$, process 121.
The difference between the two is that in process 3 the
$\t$ and $\tbar$ are added by the initial-state shower, while in
121 the full matrix element is used. The price to be paid is that
the complicated multi-body phase space in process 121 makes the
program run slower than with most other processes. As usual, it
would be double-counting to include the same flavour both with
3 and 121. An intermediate step --- in practice probably not so
useful --- is offered by process 32, $\q \g \to \q \hrm^0$,
where the quark is assumed to be a $\b$ one, with the antiquark
added by the showering activity.
Process 122 is similar in structure to 121, but is less
important. In both process 121 and 122 the produced quark is assumed
to be a $\t$; this can be changed in \ttt{KFPR(121,2)} and
\ttt{KFPR(122,2)} before initialization, however. For $\b$ quarks it
could well be that process 3 with $\b \bbar \to \hrm^0$ is more
reliable than process 121 with $\g \g \to \b \bbar \hrm^0$
\cite{Car00}; see the discussion on $\Z^0 \b \bbar$ final states in
section \ref{sss:WZclass}. Thus it would make sense to run with all
quarks up to and including $\b$ simulated in process 3 and then
consider $\t$ quarks separately in process 121. Assuming no $\t$
parton densities, this would actually be the default behaviour,
meaning that the two could be combined in the same run without
double-counting.
The two subprocess 112 and 113, with a Higgs recoiling against a
quark or gluon jet, are also effectively generated by initial-state
corrections to subprocess 102. Thus, in order to avoid double-counting,
just as for the case of $\Z^0/\W^+$ production, section \ref{sss:WZclass},
these subprocesses should not be switched on simultaneously. Process 111,
$\q\qbar \to \g\hrm^0$ is different, in the sense that it proceeds
through an $s$-channel gluon coupling to a heavy-quark loop, and that
therefore the emitted gluon is necessary in the final state in order
to conserve colours. It is not to be confused with a gluon-radiation
correction to the Born-level process 3, like in process 32, since
processes 3 and 32 vanish for massless quarks while process 111 is mainly
intended for such. The lack of a matching Born-level process shows up
by process 111 being vanishing in the $\pT \to 0$ limit. Numerically it
is of negligible importance, except at very large $\pT$ values.
Process 102, possibly augmented by 111, should thus be used for
inclusive production of Higgs, and 111--113 for the study
of the Higgs subsample with high transverse momentum.
A warning is that the matrix-element expressions for processes 111--113
are very lengthy and the coding therefore more likely to contain some
errors and numerical instabilities than for most other processes.
Therefore the full expressions are only available by setting the
non-default value \ttt{MSTP(38) = 0}. Instead the default is based on
the simplified expressions obtainable if only the top quark contribution
is considered, in the $m_{\t} \to \infty$ limit \cite{Ell88}. As a slight
improvement, this expression is rescaled by the ratio of the
$\g\g \to \hrm^0$ cross sections (or, equivalently, the $\hrm \to \g\g$
partial widths) of the full calculation and that in the
$m_{\t} \to \infty$ limit. Simple checks show that this approach normally
agrees with the full expressions to within $\sim 20$\%, which is small
compared with other uncertainties. The agreement is worse for process
111 alone, about a factor of 2, but this process is small anyway.
We also note that the matrix element correction factors, used in the
initial-state parton shower for process 102, section
\ref{sss:newinshow}, are based on the same $m_{\t} \to \infty$ limit
expressions, so that the high-$\pT$ tail of process 102 is well matched
to the simple description of process 112 and 113.
In $\ee$ annihilation, associated production of an $\hrm^0$ with a
$\Z^0$, process 24, is usually the dominant one close to threshold,
while the $\Z^0 \Z^0$ and $\W^+ \W^-$ fusion processes 123 and 124
win out at high energies. Process 103, $\gamma\gamma$ fusion, may
also be of interest, in particular when the possibilities of
beamstrahlung photons and backscattered photons are included
(see section \ref{sss:estructfun}).
Process 110, which gives an $\hrm^0$ in association with a $\gamma$,
is a loop process and is therefore suppressed in rate. It would
have been of interest for a $\hrm^0$ mass above 60 GeV at LEP 1,
since its phase space suppression there is less severe than for the
associated production with a $\Z^0$. Now it is not likely to be of
any further interest.
The branching ratios of the Higgs are very strongly dependent on
the mass. In principle, the program is set up to calculate these
correctly, as a function of the actual Higgs mass, i.e.\ not just
at the nominal mass. However, higher-order corrections may at
times be important and not fully unambiguous; see for instance
\ttt{MSTP(37)}.
Since the Higgs is a spin-0 particle it decays isotropically. In decay
processes such as $\hrm^0 \to \W^+ \W^- / \Z^0 \Z^0 \to 4$ fermions angular
correlations are included \cite{Lin97}. Also in processes 24 and 26,
$\Z^0$ and $\W^{\pm}$ decay angular distributions are correctly taken into
account.
\subsubsection{Heavy Standard Model Higgs}
\label{sss:heavySMHclass}
{\ISUB} =
\begin{tabular}[t]{rl}
5 & $\Z^0 \Z^0 \to \hrm^0$ \\
8 & $\W^+ \W^- \to \hrm^0$ \\
71 & $\Z^0 \Z^0 \to \Z^0 \Z^0$ (longitudinal) \\
72 & $\Z^0 \Z^0 \to \W^+ \W^-$ (longitudinal) \\
73 & $\Z^0 \W^+ \to \Z^0 \W^+$ (longitudinal) \\
76 & $\W^+ \W^- \to \Z^0 \Z^0$ (longitudinal) \\
77 & $\W^+ \W^{\pm} \to \W^+ \W^{\pm}$ (longitudinal) \\
\end{tabular}
Processes 5 and 8 are the simple $2 \to 1$ versions of what is now
available in 123 and 124 with the full $2 \to 3$ kinematics.
For low Higgs masses processes 5 and 8 overestimate the correct
cross sections and should not be used, whereas good agreement between
the $2 \to 1$ and $2 \to 3$ descriptions is observed when heavy
Higgs production is studied.
The subprocesses 5 and 8, $V V \to \hrm^0$, which contribute to the
processes $V V \to V' V'$, show a bad high-energy behaviour. Here
$V$ denotes a longitudinal intermediate gauge boson, $\Z^0$ or
$\W^{\pm}$. This can be cured only by the inclusion of all
$V V \to V' V'$ graphs, as is done in subprocesses 71, 72, 73, 76
and 77. In particular, subprocesses 5 and 8 give rise to a fictitious
high-mass tail of the Higgs. If this tail is thrown away, however,
the agreement between the $s$-channel graphs only (subprocesses
5 and 8) and the full set of graphs (subprocesses 71 etc.) is very
good: for a Higgs of nominal mass 300 (800) GeV, a cut at 600 (1200)
GeV retains 95\% (84\%) of the total cross section, and differs from
the exact calculation, cut at the same values, by only 2\% (11\%)
(numbers for SSC energies). With this prescription there is
therefore no need to use subprocesses 71 etc. rather than
subprocesses 5 and 8.
For subprocess 77, there is an option, see \ttt{MSTP(45)}, to select
the charge combination of the scattering $\W$'s: like-sign,
opposite-sign (relevant for Higgs), or both.
Process 77 contains a divergence for $\pT \to 0$ due to
$\gamma$-exchange contributions. This leads to an infinite total
cross section, which is entirely fictitious, since the simple
parton-distribution function approach to the longitudinal $\W$ flux
is not appropriate in this limit. For this process, it is therefore
necessary to make use of a cut, e.g.\ $\pT > m_{\W}$.
For subprocesses 71, 72, 76 and 77, an option is included (see
\ttt{MSTP(46)}) whereby you can select only the $s$-channel
Higgs graph; this will then be essentially equivalent to running
subprocess 5 or 8 with the proper decay channels (i.e.\ $\Z^0\Z^0$ or
$\W^+\W^-$) set via \ttt{MDME}. The difference is that the
Breit--Wigner distributions in subprocesses 5 and 8 contain a mass-dependent
width, whereas the width in subprocesses 71--77 is calculated at
the nominal Higgs mass; also, higher-order corrections to the widths
are treated more accurately in subprocesses 5 and 8. Further,
processes 71--77 assume the incoming $\W/\Z$ to be on the mass shell,
with associated kinematics factors, while processes 5 and 8 have
$\W/\Z$ correctly space-like. All this leads to differences in the
cross sections by up to a factor of 1.5.
In the absence of a Higgs, the sector of longitudinal $\Z$ and $\W$
scattering will become strongly interacting at energies above 1 TeV.
The models proposed by Dobado, Herrero and Terron \cite{Dob91} to
describe this kind of physics have been included as alternative matrix
elements for subprocesses 71, 72, 73, 76 and 77, selectable by
\ttt{MSTP(46)}. From the point of view of the general classification
scheme for subprocesses, this model should appropriately be
included as separate subprocesses with numbers above 100, but the
current solution allows a more efficient reuse of existing code.
By a proper choice of parameters, it is also here possible to
simulate the production of a techni-$\rho$ (see section
\ref{sss:technicolorclass}).
Currently, the scattering of transverse gauge bosons has not been
included, neither that of mixed transverse--longitudinal scatterings.
These are expected to be less important at high energies, and do not
contain an $\hrm^0$ resonance peak, but need not be
entirely negligible in magnitude. As a rule of thumb, processes
71--77 should not be used for $VV$ invariant masses below 500 GeV.
The decay products of the longitudinal gauge bosons are correctly
distributed in angle.
\subsubsection{Extended neutral Higgs sector}
\label{sss:extneutHclass}
\ttt{MSEL} = 19 \\
{\ISUB} =
\begin{tabular}[t]{rrrl}
$\hrm^0$ & $\H^0$ & $\A^0$ & \\
3 & 151 & 156 & $\f_i \fbar_i \to X$ \\
102 & 152 & 157 & $\g \g \to X$ \\
103 & 153 & 158 & $\gamma \gamma \to X$ \\
111 & 183 & 188 & $\q \qbar \to \g X$ \\
112 & 184 & 189 & $\q \g \to \q X$ \\
113 & 185 & 190 & $\g \g \to \g X$ \\
24 & 171 & 176 & $\f_i \fbar_i \to \Z^0 X$ \\
26 & 172 & 177 & $\f_i \fbar_j \to \W^+ X$ \\
123 & 173 & 178 & $\f_i \f_j \to \f_i \f_j X$ ($\Z \Z$ fusion) \\
124 & 174 & 179 & $\f_i \f_j \to \f_k \f_l X$ ($\W^+\W^-$ fusion) \\
121 & 181 & 186 & $\g \g \to \Q_k \Qbar_k X$ \\
122 & 182 & 187 & $\q_i \qbar_i \to \Q_k \Qbar_k X$ \\
\end{tabular}
In {\Py}, the particle content of a two-Higgs-doublet scenario is
included: two neutral scalar particles, 25 and 35, one pseudoscalar
one, 36, and a charged doublet, $\pm 37$. (Of course, these particles
may also be associated with corresponding Higgs states in larger
multiplets.) By convention, we choose to call the lighter scalar
Higgs $\hrm^0$ and the heavier $\H^0$. The pseudoscalar is called
$\A^0$ and the charged $\H^{\pm}$. Charged-Higgs production is covered
in section \ref{sss:chHclass}.
A number of $\hrm^0$ processes have been duplicated for $\H^0$ and
$\A^0$. The correspondence between {\ISUB} numbers is shown in the table
above: the first column of {\ISUB} numbers corresponds to
$X = \hrm^0$, the second to $X = \H^0$, and the third to $X = \A^0$.
Note that several of these processes are not expected to take
place at all, owing to vanishing Born term couplings. We have still
included them for flexibility in simulating arbitrary couplings at
the Born or loop level, or for the case of mixing between the
scalar and pseudoscalar sectors.
A few Standard Model Higgs processes have no correspondence in the
scheme above. These include
\begin{Itemize}
\item 5 and 8, which anyway have been superseded by 123 and 124;
\item 71, 72, 73, 76 and 77, which deal with what happens if there
is no light Higgs, and so is a scenario complementary to the one
above, where several light Higgs bosons are assumed; and
\item 110, which is mainly of interest in Standard Model Higgs
searches.
\end{Itemize}
The processes 102--103, 111--113, 152--153, 157--158,
183--185 and 188--190 have only been worked
out in full detail for the Standard Model Higgs case, and not when
other (e.g. squark loop) contributions need be considered.
For processes 102--103, 152--153, and 157--158, the same approximation
mainly holds true
for the decays, since these production processes are proportional
to the partial decay width for the $\g\g$ and $\gamma\gamma$ channels.
The $\gamma\gamma$ channel does include $\H^+$ in the loop.
In some corners of SUSY parameter space, the effects of squarks and
gauginos in loops can be relevant.
The approximate
procedure outlined in section \ref{sss:lightSMHclass}, based on
combining the kinematics shape from simple expressions in the
$m_{\t} \to \infty$ limit with a normalization derived from the
$\g\g \to X$ cross section, should therefore be viewed as a first
ansatz only. In particular, it is not recommended to try the
non-default \ttt{MSTP(38) = 0} option, which is incorrect beyond the
Standard Model.
In processes 121, 122, 181, 182, 186 and 187 the recoiling heavy
flavour is assumed to be top, which is the only one of interest in
the Standard Model, and the one where the parton-distribution-function
approach invoked in processes 3, 151 and 156 is least reliable.
However, it is possible to change the quark flavour in 121 etc.;
for each process {\ISUB} this flavour is given by \ttt{KFPR(ISUB,2)}.
This may become relevant if couplings to $\b\bbar$ states are
enhanced, e.g.\ if $\tan\beta \gg 1$ in the MSSM. The matrix elements
in this group are based on scalar Higgs couplings; differences for
a pseudoscalar Higgs remains to be worked out, but are proportional
to the heavy quark mass relative to other kinematic quantities.
By default, the $\hrm^0$ has the couplings of the Standard Model
Higgs, while the $\H^0$ and $\A^0$ have couplings set in
\ttt{PARU(171) - PARU(178)} and \ttt{PARU(181) - PARU(190)},
respectively. The default values for the $\H^0$ and $\A^0$ have no
deep physics motivation, but are set just so that the program will
not crash due to the absence of any couplings whatsoever. You
should therefore set the above couplings to your desired values if
you want to simulate either $\H^0$ or $\A^0$. Also the couplings
of the $\hrm^0$ particle can be modified, in
\ttt{PARU(161) - PARU(165)}, provided that \ttt{MSTP(4) = 1}.
For \ttt{MSTP(4) = 2}, the mass of the $\hrm^0$ (\ttt{PMAS(25,1)})
and the $\tan\beta$ value (\ttt{PARU(141)}) are used to derive
the masses of the other Higgs bosons, as well as all Higgs couplings.
\ttt{PMAS(35,1) - PMAS(37,1)} and \ttt{PARU(161) - PARU(195)} are
overwritten accordingly. The relations used are the ones of the
Born-level MSSM \cite{Gun90}.
Note that not all combinations of $m_{\hrm}$ and $\tan\beta$ are
allowed; for \ttt{MSTP(4) = 2} the requirement of a finite $\A^0$ mass
imposes the constraint
\begin{equation}
m_{\hrm} < m_{\Z} \, \frac{\tan^2\beta - 1}{\tan^2\beta + 1},
\end{equation}
or, equivalently,
\begin{equation}
\tan^2\beta > \frac{m_{\Z} + m_{\hrm}}{m_{\Z} - m_{\hrm}}.
\end{equation}
If this condition is not fulfilled, the program will print
a diagnostic message and stop.
A more realistic approach to the Higgs mass spectrum is to
include radiative corrections to the Higgs potential. Such a
machinery has never been implemented as such in {\Py}, but
appears as part of the Supersymmetry framework described in
sections \ref{ss:susyclass} and \ref{ss:susycode}. At tree level,
the minimal set of inputs would be \ttt{IMSS(1) = 1} to switch on
SUSY, \ttt{RMSS(5)} to set the $\tan\beta$ value
(this overwrites the \ttt{PARU(141)} value when SUSY is
switched on) and \ttt{RMSS(19)} to set $\A^0$ mass.
However, the significant radiative corrections depend
on the properties of all particles that couple to the
Higgs boson, and the user may want to change the default values
of the relevant \ttt{RMSS} inputs. In practice, the most
important are those related indirectly to the physical masses of
the third generation supersymmetric quarks and the Higgsino:
\ttt{RMSS(10)} to set the left-handed doublet SUSY mass
parameter, \ttt{RMSS(11)} to set the right stop mass parameter,
\ttt{RMSS(12)} to set the right sbottom mass parameter,
\ttt{RMSS(4)} to set the Higgsino mass and a portion of the
squark mixing, and \ttt{RMSS(16)} and \ttt{RMSS(17)} to set the
stop and bottom trilinear couplings, respectively, which
specifies the remainder of the squark mixing. From these inputs,
the Higgs masses and couplings would be derived. Note that
switching on SUSY also implies that Supersymmetric decays
of the Higgs particles become possible if kinematically allowed.
If you do not want this to happen, you may want to increase the
SUSY mass parameters. (Use \ttt{CALL PYSTAT(2)} after
initialization to see the list of branching ratios.)
Pair production of Higgs states may be a relevant source, see
section \ref{sss:Hpairclass} below.
Finally, heavier Higgs bosons may decay into lighter ones, if
kinematically allowed, in processes like $\A^0 \to \Z^0 \hrm^0$ or
$\H^+ \to \W^+ \hrm^0$. Such modes are included as part of the
general mixture of decay channels, but they can be enhanced if
the uninteresting channels are switched off.
\subsubsection{Charged Higgs sector}
\label{sss:chHclass}
\ttt{MSEL} = 23 \\
{\ISUB} =
\begin{tabular}[t]{rl}
143 & $\f_i \fbar_j \to \H^+$ \\
161 & $\f_i \g \to \f_k \H^+$ \\
401 & $\g \g \to \tbar \b \H^+$ \\
402 & $\q \qbar \to \tbar \b \H^+$ \\
\end{tabular}
A charged Higgs doublet, $\H^{\pm}$, is included in the program.
This doublet may be the one predicted in the MSSM scenario,
see section \ref{sss:extneutHclass}, or in any other scenario.
The $\tan\beta$ parameter, which is relevant also
for neutral Higgs couplings, is set via \ttt{PARU(141)} or,
if SUSY is switched on, via \ttt{RMSS(5)}.
The basic subprocess for charged Higgs production in hadron
colliders is {\ISUB} = 143. However, this process is dominated
by $\t \bbar \to \H^+$, and so depends on the choice of $\t$
parton distribution, if non-vanishing. A better representation
is provided by subprocess 161, $\f \g \to \f' \H^+$; i.e.\ actually
$\bbar \g \to \tbar \H^+$. It is therefore recommended to use
161 and not 143; to use both would be double-counting.
A further step is to include the initial state gluon splitting
$\g \to \b \bbar$ as part of the matrix element,
as in process 401. Using both 161 and 401 again
would involve doublecounting, but now the issue is more complicated,
since 401 may be expected to give the better description at large
$p_{\perp\b}$ and 161 the better at small $p_{\perp\b}$, so some
matching may be the best solution \cite{Bor99}. Process 402 gives
a less important contribution, that can be added without
doublecounting. (The situation is similar to that for processes
3, 32, 121 and 122 for neutral Higgs production.)
A major potential source of charged Higgs production is top decay.
It is possible to switch on the decay channel $\t \to \b \H^+$. Top
will then decay to $\H^+$ a fraction of the time, whichever way it is
produced. The branching ratio is automatically calculated, based
on the $\tan\beta$ value and masses.
It is possible to only have the $\H^+$ decay mode switched on,
in which case the cross section is reduced accordingly.
Pair production of the charged Higgs is also possible through
its electromagnetic charge; see
section \ref{sss:Hpairclass} below.
\subsubsection{Higgs pairs}
\label{sss:Hpairclass}
{\ISUB} =
\begin{tabular}[t]{rl}
(141) & $\f_i \fbar_i \to \gamma/\Z^0/\Z'^0$ \\
297 & $\f_i \fbar_j \to \H^{\pm} \hrm^0$ \\
298 & $\f_i \fbar_j \to \H^{\pm} \H^0$ \\
299 & $\f_i \fbar_i \to \A \hrm^0$ \\
300 & $\f_i \fbar_i \to \A \H^0$ \\
301 & $\f_i \fbar_i \to \H^+ \H^-$ \\
\end{tabular}
The subprocesses 297--301 give the production of a pair of Higgs
bosons via the $s$-channel exchange of a
$\gamma^*/\Z^0$ or a $\W^{\pm}$ state.
Note that Higgs pair production is possible alternatively through
subprocess 141, as part of the decay of a generic combination of
$\gamma^*/\Z^0/\Z'^0$. Thus it can be used to simulate
$\Z^0 \to \hrm^0 \A^0$ and $\Z^0 \to \H^0 \A^0$ for associated
neutral Higgs production. The fact that we here make use of the
$\Z'^0$ can easily be discounted, either by letting the relevant
couplings vanish, or by the option \ttt{MSTP(44) = 4}.
Similarly the decay $\gamma^*/\Z^0/\Z'^0 \to \H^+ \H^-$
allows the production of a pair of charged Higgs particles.
This process is especially important in $\ee$ colliders.
The coupling of the $\gamma^*$ to $\H^+ \H^-$ is determined by
the charge alone (neglecting loop effects), while the $\Z^0$
coupling is regulated by \ttt{PARU(142)}, and that of the
$\Z'^0$ by \ttt{PARU(143)}. The $\Z'^0$ piece can be switched off,
e.g.\ by \ttt{MSTP(44) = 4}. An ordinary $\Z^0$, i.e.\ particle code 23,
cannot be made to decay into a Higgs pair, however.
The advantage of the explicit pair production processes is the
correct implementation of the pair threshold.
\subsection{Non-Standard Physics}
Many extensions of the Standard Model have been proposed, and likely
this represents only a small number of what is possible.
Despite different underlying assumptions, many Beyond-the-Standard-Model
scenarios predict some common interactions or particles.
For instance, new $\W'$ and $\Z'$ gauge bosons can
arise in a number of different ways. Therefore it makes sense
to cover a few basic classes of particles and interactions, with enough
generality that many kinds of detailed scenarios can be
accommodated by suitable parameter choices. We have already seen one
example of this, in the extended Higgs sector above. In this section
a few other kinds of non-standard generic physics scenarios are discussed.
Supersymmetry is covered separately in the following section,
since it is such a large sector by itself.
\subsubsection{Fourth-generation fermions}
\ttt{MSEL} = 7, 8, 37, 38 \\
{\ISUB} =
\begin{tabular}[t]{rl}
1 & $\f_i \fbar_i \to \gammaZ$ \\
2 & $\f_i \fbar_j \to \W^+$ \\
81 & $\q_i \qbar_i \to \Q_k \Qbar_k$ \\
82 & $\g \g \to \Q_k \Qbar_k$ \\
83 & $\q_i \f_j \to \Q_k \f_l$ \\
84 & $\g \gamma \to \Q_k \Qbar_k$ \\
85 & $\gamma \gamma \to \F_k \Fbar_k$ \\
141 & $\f_i \fbar_i \to \gamma/\Z^0/\Z'^0$ \\
142 & $\f_i \fbar_j \to \W'^+$ \\
\end{tabular}
While the existence of a
fourth generation currently seems unlikely,
the appropriate flavour content is still found in the program. In
fact, the fourth generation is included on an equal basis with the
first three, provided \ttt{MSTP(1) = 4}. Also processes other than
the ones above can therefore be used, e.g.\ all other processes with
gauge bosons, including non-standard ones such as the $\Z'^0$. We
therefore do not repeat the descriptions found elsewhere, such as how
to set only the desired flavour in processes 81--85. Note that it
may be convenient to set \ttt{CKIN(1)} and other cuts such that the
mass of produced gauge bosons is enough for the wanted particle
production --- in principle the program will cope even without that,
but possibly at the expense of very slow execution.
\subsubsection{New gauge bosons}
\ttt{MSEL} = 21, 22, 24 \\
{\ISUB} =
\begin{tabular}[t]{rl}
141 & $\f_i \fbar_i \to \gamma/\Z^0/\Z'^0$ \\
142 & $\f_i \fbar_j \to \W'^+$ \\
144 & $\f_i \fbar_j \to \R$ \\
\end{tabular}
The $\Z'^0$ of subprocess 141 contains the full $\gamma^*/\Z^0/\Z'^0$
interference structure for couplings to fermion pairs. With
\ttt{MSTP(44)} it is possible to pick only a subset, e.g.\ only the
pure $\Z'^0$ piece. The couplings of the $\Z'^0$ to quarks and leptons
in the first generation can be set via \ttt{PARU(121) - PARU(128)},
in the second via \ttt{PARJ(180) - PARJ(187)} and in the third via
\ttt{PARJ(188) - PARJ(195)}. The eight numbers
correspond to the vector and axial couplings of down-type quarks,
up-type quarks, leptons and neutrinos, respectively. The default
corresponds to the same couplings as that of the Standard Model
$\Z^0$, with axial couplings $a_{\f} = \pm 1$ and vector couplings
$v_{\f} = a_{\f} - 4 e_{\f} \ssintw$. This implies a resonance width
that increases linearly with the mass. By a suitable choice of the
parameters, it is possible to simulate just about any imaginable
$\Z'^0$ scenario, with full interference effects in cross sections
and decay angular distributions and generation-dependent
couplings.
The conversion from the coupling conventions in a set of different
$\Z'^0$ models in the literature
to those used in {\Py} can be found in \cite{Cio05}.
The coupling to the decay channel $\Z'^0 \to \W^+ \W^-$ is regulated
by \ttt{PARU(129) - PARU(130)}. The former gives the strength of the
coupling, which determines the rate. The default, \ttt{PARU(129) = 1.},
corresponds to the `extended gauge model' of \cite{Alt89}, wherein
the $\Z^0 \to \W^+ \W^-$ coupling is used, scaled down by a factor
$m_{\W}^2/m_{\Z'}^2$, to give a $\Z'^0$ partial width into this channel
that again increases linearly. If this factor is cancelled, by having
\ttt{PARU(129)} proportional to $m_{\Z'}^2/m_{\W}^2$, one obtains a
partial width that goes like the fifth power of the $\Z'^0$ mass, the
`reference model' of \cite{Alt89}. In the decay angular distribution
one could imagine a much richer structure than is given by the one
parameter \ttt{PARU(130)}.
Other decay modes include $\Z'^0 \to \Z^0 \hrm^0$, predicted in
left--right symmetric models (see \ttt{PARU(145)} and ref.
\cite{Coc91}), and a number of other Higgs decay channels, see
sections \ref{sss:extneutHclass} and \ref{sss:chHclass}.
The $\W'^{\pm}$ of subprocess 142 so far does not contain
interference with the Standard Model $\W^{\pm}$ --- in practice this
should not be a major limitation. The couplings of the $\W'$ to
quarks and leptons are set via \ttt{PARU(131) - PARU(134)}.
Again one may set vector and axial couplings freely, separately
for the $\q \qbar'$ and the $\ell \nu_{\ell}$ decay channels.
The defaults correspond to the $V-A$ structure of the Standard Model
$\W$, but can be changed to simulate a wide selection of models.
One possible limitation is that the same Cabibbo--Kobayashi--Maskawa
quark mixing matrix is assumed as for the standard $\W$.
The coupling $\W' \to \Z^0 \W$ can be set via
\ttt{PARU(135) - PARU(136)}. Further comments on this channel as for
$\Z'$; in particular, default couplings again agree with
the `extended gauge model' of \cite{Alt89}. A $\W' \to \W \hrm^0$
channel is also included, in analogy with the $\Z'^0 \to \Z^0 \hrm^0$
one, see \ttt{PARU(146)}.
The $\R$ boson (particle code 41) of subprocess 144 represents one
possible scenario \cite{Ben85a} for a horizontal gauge boson,
i.e.\ a gauge boson
that couples between the generations, inducing processes like
$\s \dbar \to \R^0 \to \mu^- \e^+$. Experimental limits on
flavour-changing neutral currents forces such a boson to be fairly
heavy.
A further example of new gauge groups follows right after this.
\subsubsection{Left--Right Symmetry and Doubly Charged Higgs Bosons}
\label{sss:LRDCHclass}
{\ISUB} =
\begin{tabular}[t]{rl}
341 & $\ell_i \ell_j \to \H_L^{\pm\pm}$ \\
342 & $\ell_i \ell_j \to \H_R^{\pm\pm}$ \\
343 & $\ell_i \gamma \to \H_L^{\pm\pm} \e^{\mp}$ \\
344 & $\ell_i \gamma \to \H_R^{\pm\pm} \e^{\mp}$ \\
345 & $\ell_i \gamma \to \H_L^{\pm\pm} \mu^{\mp}$ \\
346 & $\ell_i \gamma \to \H_R^{\pm\pm} \mu^{\mp}$ \\
347 & $\ell_i \gamma \to \H_L^{\pm\pm} \tau^{\mp}$ \\
348 & $\ell_i \gamma \to \H_R^{\pm\pm} \tau^{\mp}$ \\
349 & $\f_i \fbar_i \to \H_L^{++} \H_L^{--}$ \\
350 & $\f_i \fbar_i \to \H_R^{++} \H_R^{--}$ \\
351 & $\f_i \f_j \to \f_k f_l \H_L^{\pm\pm}$ ($\W\W$ fusion) \\
352 & $\f_i \f_j \to \f_k f_l \H_R^{\pm\pm}$ ($\W\W$ fusion) \\
353 & $\f_i \fbar_i \to \Z_R^0$ \\
354 & $\f_i \fbar_i \to \W_R^{\pm}$ \\
\end{tabular}
At current energies, the world is left-handed, i.e.\ the Standard Model
contains an {\bf SU(2)}$_L$ group. Left--right symmetry at
some larger scale implies the need for an {\bf SU(2)}$_R$ group.
Thus the particle content is expanded by right-handed $\Z_R^0$ and
$\W_R^{\pm}$ and right-handed neutrinos. The Higgs fields have to be
in a triplet representation, leading to doubly-charged Higgs particles,
one set for each of the two {\bf SU(2)} groups. Also the number of
neutral and singly-charged Higgs states is increased relative to the
Standard Model, but a search for the lowest-lying states of this kind
is no different from e.g.\ the freedom already accorded by the MSSM
Higgs scenarios.
{\Py} implements the scenario of \cite{Hui97}. The expanded particle
content with default masses is:\\
\begin{tabular}{llr}
{\KF} & name & $m$ (GeV)\\
9900012 & $\nu_{R\e}$ & 500 \\
9900014 & $\nu_{R\mu}$ & 500 \\
9900016 & $\nu_{R\tau}$& 500 \\
9900023 & $\Z_R^0$ & 1200 \\
9900024 & $\W_R^+$ & 750 \\
9900041 & $\H_L^{++}$ & 200 \\
9900042 & $\H_R^{++}$ & 200 \\
\end{tabular}\\
The main decay modes implemented are
$\H_L^{++} \to \W_L^+ \W_L^+, \ell_i^+ \ell_j^+$ ($i, j$ generation
indices) and
$\H_R^{++} \to \W_R^+ \W_R^+, \ell_i^+ \ell_j^+$.
The physics parameters of the scenario are found in
\ttt{PARP(181) - PARP(192)}.
The $W_R^{\pm}$ has been implemented as a simple copy of the
ordinary $\W^{\pm}$, with the exception that it couple to
right-handed neutrinos instead of the ordinary left-handed ones.
Thus the standard CKM matrix is used in the quark sector, and the
same vector and axial coupling strengths, leaving only the mass as
free parameter. The $\Z_R^0$ implementation (without interference
with $\gamma$ or the ordinary $\Z^0$) allows decays both to left-
and right-handed neutrinos, as well as other fermions, according to
one specific model ansatz \cite{Fer00}. Obviously both the $W_R^{\pm}$
and the $\Z_R^0$ descriptions are likely to be simplifications,
but provide a starting point.
The right-handed neutrinos can be allowed to decay further
\cite{Riz81,Fer00}. Assuming them to have a mass below that of
$\W_R^+$, they decay to three-body states via a virtual $\W_R^+$,
$\nu_{R\ell} \to \ell^+ \f \fbar'$ and
$\nu_{R\ell} \to \ell^- \fbar \f'$, where both choices are allowed
owing to the Majorana character of the neutrinos. If there is
a significant mass splitting, also sequential decays
$\nu_{R\ell} \to \ell^{\pm} {\ell'}^{\mp} {\nu'}_{R\ell}$
are allowed. Currently the decays are isotropic in phase space.
If the neutrino masses are close to or above the $\W_R$ ones, this
description has to be substituted by a sequential decay via
a real $\W_R$ (not implemented, but actually simpler to do than the
one here).
\subsubsection{Leptoquarks}
\label{sss:LQclass}
\ttt{MSEL} = 25 \\
{\ISUB} =
\begin{tabular}[t]{rl}
145 & $\q_i \ell_j \to \L_{\Q}$ \\
162 & $\q \g \to \ell \L_{\Q}$ \\
163 & $\g \g \to \L_{\Q} \br{\L}_{\Q}$ \\
164 & $\q_i \qbar_i \to \L_{\Q} \br{\L}_{\Q}$ \\
\end{tabular}
Several leptoquark production processes are included.
Currently only one leptoquark species has been implemented, as particle 42,
denoted $\L_{\Q}$. The leptoquark is assumed to carry specific quark
and lepton quantum numbers, by default $\u$ quark plus electron.
These flavour numbers are conserved, i.e.\ a process such as
$\u \e^- \to \L_{\Q} \to \d \nu_{\e}$ is not allowed. This may be a
bit restrictive, but it represents one of many leptoquark possibilities.
The spin of the leptoquark is assumed to be zero, so its decay is
isotropic. Vector leptoquarks have not yet been implemented.
Although only one leptoquark is implemented, its flavours may be
changed arbitrarily to study the different possibilities. The
flavours of the leptoquark are defined by the quark and lepton
flavours in the decay mode list. Since only one decay channel is
allowed, this means that the quark flavour is stored in
\ttt{KFDP(MDCY(42,2),1)} and the lepton one in
\ttt{KFDP(MDCY(42,2),2)}. The former must always be a quark, while
the latter could be a lepton or an antilepton; a charge-conjugate
partner is automatically defined by the program. At initialization,
the charge is recalculated as a function of the flavours defined;
also the leptoquark name is redefined to be of the type
\ttt{'LQ\_(q)(l)'}, where actual quark \ttt{(q)} and lepton \ttt{(l)}
flavours are displayed.
The $\L_{\Q} \to \q \ell$ vertex contains an undetermined Yukawa
coupling strength, which fixes both the width of the leptoquark
and the cross section for many of the production graphs. This
strength may be changed in \ttt{PARU(151)} which
corresponds to the $k$ factor of \cite{Hew88}, i.e.\
to $\lambda^2/(4\pi\alphaem)$, where $\lambda$ is the Yukawa
coupling strength of \cite{Wud86}. Note that \ttt{PARU(151)}
is thus quadratic in the coupling.
The leptoquark is likely to be fairly long-lived, in which case it
has time to fragment into a mesonic- or baryonic-type state, which
would decay later on. This is a bit tedious to handle; therefore
the leptoquark is always assumed to decay before fragmentation.
This simplification should not have a major impact
on experimental analyses \cite{Fri97}.
Inside the program, the leptoquark is treated as a resonance.
Since it carries colour, some extra care is required.
In particular, the leptoquark should not be made stable
by modifying either \ttt{MDCY(42,1)} or \ttt{MSTP(41)}: then the
leptoquark would be handed undecayed to {\Py}, which would try
to fragment it (as it does with any other coloured object)
unsuccessfully, leading to error messages and a premature
conclusion of the run.
\subsubsection{Compositeness and anomalous couplings}
\label{sss:ancoupclass}
{\ISUB} =
\begin{tabular}[t]{rl}
20 & $\f_i \fbar_j \to \gamma \W^+$ \\
165 & $\f_i \fbar_i \to \f_k \fbar_k$ (via $\gammaZ$) \\
166 & $\f_i \fbar_j \to \f_k \fbar_l$ (via $\W^{\pm}$) \\
381 & $\f_i \f_j \to \f_i \f_j$ \\
382 & $\f_i \fbar_i \to \f_k \fbar_k$ \\
\end{tabular}
Some processes are implemented to allow the introduction of
anomalous coupling,
in addition to the Standard Model ones. These can be
switched on by \ttt{ITCM(5)} $\geq 1$; the default \ttt{ITCM(5) = 0}
corresponds to the Standard Model behaviour.
In processes 381 and 382, the quark substructure is included in the
left--left isoscalar model \cite{Eic84,Chi90} for
\ttt{ITCM(5) = 1}, with compositeness
scale $\Lambda$ given in \ttt{RTCM(41)} (default 1000~GeV) and sign
$\eta$ of interference term in \ttt{RTCM(42)} (default $+1$; only
other alternative $-1$). The above model assumes that only $\u$ and
$\d$ quarks are composite (at least at the scale studied); with
\ttt{ITCM(5) = 2} compositeness terms are included in the
interactions between all quarks. When \ttt{ITCM(5) = 0}, the two
processes are equivalent with 11 and 12. A consistent set of high-$\pT$
jet production processes in compositeness scenarios is thus obtained
by combining 381 and 382 with 13, 28, 53 and 68.
The processes 165 and 166 are basically equivalent to 1 and 2, i.e.\
$\gammaZ$ and $\W^{\pm}$ exchange, respectively, but with less detail
(no mass-dependent width, etc.). The reason for this duplication
is that the resonance treatment formalism of processes 1 and 2 could
not easily be extended to include other than $s$-channel graphs.
In processes 165 and 166, only one final-state flavour
is generated at the time; this flavour should be set in
\ttt{KFPR(165,1)} and \ttt{KFPR(166,1)}, respectively. For process
166 one gives the down-type flavour, and the program will associate
the up-type flavour of the same generation. Defaults are 11 in both
cases, i.e.\ $\ee$ and $\e^+ \nu_{\e}$ ($\e^- \br{\nu}_{\e}$) final
states. While \ttt{ITCM(5) = 0} gives the Standard Model results,
\ttt{ITCM(5) = 1} contains the left--left isoscalar model (which
does not affect process 166), and \ttt{ITCM(5) = 3} the
helicity-non-conserving model (which affects both) \cite{Eic84,Lan91}.
Both models above assume that only $\u$ and $\d$ quarks are
composite; with \ttt{ITCM(5) =} 2 or 4, respectively, contact terms
are included for all quarks in the initial state. The relevant
parameters are
\ttt{RTCM(41)} and \ttt{RTCM(42)}, as above.
Note that processes 165 and 166 are book-kept as $2 \to 2$ processes,
while 1 and 2 are $2 \to 1$ ones. This means that the default $\Q^2$
scale in parton distributions is $\pT^2$ for the former and $\hat{s}$
for the latter. To make contact between the two, it is recommended to
set \ttt{MSTP(32) = 4}, so as to use $\hat{s}$ as scale also for
processes 165 and 166.
In process 20, for $\W \gamma$ pair production, it is possible to set
an anomalous magnetic moment for the $\W$ in \ttt{RTCM(46)}
($= \eta = \kappa-1$; where $\kappa = 1$ is the Standard Model
value). The production process is affected according to the formulae
of \cite{Sam91}, while $\W$ decay currently remains
unaffected. It is necessary to set \ttt{ITCM(5) = 1} to enable this
extension.
\subsubsection{Excited fermions}
\label{sss:qlstarclass}
{\ISUB} =
\begin{tabular}[t]{rl}
146 & $\e \gamma \to \e^*$ \\
147 & $\d \g \to \d^*$ \\
148 & $\u \g \to \u^*$ \\
167 & $\q_i \q_j \to \q_k \d^*$ \\
168 & $\q_i \q_j \to \q_k \u^*$ \\
169 & $\q_i \qbar_i \to \e^{\pm} \e^{*\mp}$ \\
\end{tabular}
Compositeness scenarios may also give rise to sharp resonances of
excited quarks and leptons. An excited copy of the first generation
is implemented, consisting of spin $1/2$ particles $\d^*$ (code 4000001),
$\u^*$ (4000002), $\e^*$ (4000011) and $\nu^*_{\e}$ (4000012).
A treatment of other generations is not currently possible.
The current implementation contains gauge interaction production
by quark--gluon fusion (processes 147 and 148) or lepton--photon fusion
(process 146) and contact interaction production by quark--quark or
quark--antiquark scattering (processes 167--169) . The couplings $f$,
$f'$ and $f_s$ to the {\bf SU(2)}, {\bf U(1)} and {\bf SU(3)} groups
are stored in \ttt{RTCM(43) - RTCM(45)}, the scale parameter
$\Lambda$ in \ttt{RTCM(41)}; you are also expected to change the
$\f^*$ masses in accordance with what is desired --- see \cite{Bau90}
for details on conventions. Decay processes are of
the types $\q^* \to \q \g$, $\q^* \to \q \gamma$,
$\q^* \to \q \Z^0$ or $\q^* \to \q' \W^{\pm}$, with the latter
three (two) available also for $\e^*$ ($\nu^*_{\e}$).
A non-trivial angular dependence is included in the $\q^*$
decay for processes 146--148, but has not been included for
processes 167--169.
\subsubsection{Technicolor}
\label{sss:technicolorclass}
\ttt{MSEL} = 50, 51 \\
{\ISUB} =
\begin{tabular}[t]{rl}
149 & $\g \g \to \eta_{\mrm{tc}}$ (obsolete) \\
191 & $\f_i \fbar_i \to \rho_{\mrm{tc}}^0$ (obsolete) \\
192 & $\f_i \fbar_j \to \rho_{\mrm{tc}}^+$ (obsolete) \\
193 & $\f_i \fbar_i \to \omega_{\mrm{tc}}^0$ (obsolete) \\
194 & $\f_i \fbar_i \to \f_k \fbar_k$ \\
195 & $\f_i \fbar_j \to \f_k \fbar_l$ \\
361 & $\f_i \fbar_i \to \W^+_{\mrm{L}} \W^-_{\mrm{L}} $ \\
362 & $\f_i \fbar_i \to \W^{\pm}_{\mrm{L}} \pi^{\mp}_{\mrm{tc}}$ \\
363 & $\f_i \fbar_i \to \pi^+_{\mrm{tc}} \pi^-_{\mrm{tc}}$ \\
364 & $\f_i \fbar_i \to \gamma \pi^0_{\mrm{tc}} $ \\
365 & $\f_i \fbar_i \to \gamma {\pi'}^0_{\mrm{tc}} $ \\
366 & $\f_i \fbar_i \to \Z^0 \pi^0_{\mrm{tc}} $ \\
367 & $\f_i \fbar_i \to \Z^0 {\pi'}^0_{\mrm{tc}} $ \\
368 & $\f_i \fbar_i \to \W^{\pm} \pi^{\mp}_{\mrm{tc}}$ \\
370 & $\f_i \fbar_j \to \W^{\pm}_{\mrm{L}} \Z^0_{\mrm{L}} $ \\
371 & $\f_i \fbar_j \to \W^{\pm}_{\mrm{L}} \pi^0_{\mrm{tc}}$ \\
372 & $\f_i \fbar_j \to \pi^{\pm}_{\mrm{tc}} \Z^0_{\mrm{L}} $ \\
373 & $\f_i \fbar_j \to \pi^{\pm}_{\mrm{tc}} \pi^0_{\mrm{tc}} $ \\
374 & $\f_i \fbar_j \to \gamma \pi^{\pm}_{\mrm{tc}} $ \\
375 & $\f_i \fbar_j \to \Z^0 \pi^{\pm}_{\mrm{tc}} $ \\
376 & $\f_i \fbar_j \to \W^{\pm} \pi^0_{\mrm{tc}} $ \\
377 & $\f_i \fbar_j \to \W^{\pm} {\pi'}^0_{\mrm{tc}}$ \\
381 & $\q_i \q_j \to \q_i \q_j$ \\
382 & $\q_i \qbar_i \to \q_k \qbar_k$ \\
383 & $\q_i \qbar_i \to \g \g$ \\
384 & $\f_i \g \to \f_i \g$ \\
385 & $\g \g \to \q_k \qbar_k$ \\
386 & $\g \g \to \g \g$ \\
387 & $\f_i \fbar_i \to \Q_k \Qbar_k$ \\
388 & $\g \g \to \Q_k \Qbar_k$ \\
\end{tabular}
Technicolor (TC) uses strong
dynamics instead of weakly-coupled fundamental scalars to
manifest the Higgs mechanism for giving masses to the $\W$ and $\Z$ bosons.
In TC, the breaking of a chiral
symmetry in a new, strongly interacting gauge theory generates the
Goldstone bosons necessary for electroweak symmetry breaking (EWSB).
Thus three of the technipions assume the r\^ole of the longitudinal
components of the $\W$ and $\Z$ bosons, but other states can remain
as separate particles depending on the gauge group:
technipions ($\pi_{\mrm{tc}}$), technirhos
($\rho_{\mrm{tc}}$), techniomegas ($\omega_{\mrm{tc}}$), etc.
No fully-realistic model of strong EWSB
has been found so far, and some of the assumptions and simplifications
used in model-building may need to be discarded in the future.
The processes represented here correspond to
several generations of development. Processes 149, 191, 192 and 193
should be considered obsolete and superseded by the other
processes 194, 195 and 361--377. The former processes are kept for
cross-checks and backward compatibility.
In section \ref{sss:heavySMHclass} it is discussed how processes
71--77 can be used to simulate a scenario
with techni-$\rho$ resonances in longitudinal gauge boson scattering.
Process 149 describes the production of
a spin--0 techni-$\eta$ meson (particle code
{\KF} = 3000331), which is an electroweak singlet and a QCD colour octet.
It is one of the possible techni-$\pi$ particles; the name
`techni-$\eta$' is not used universally in the literature.
The techni-$\eta$ couples to ordinary fermions proportional to fermion
mass. The dominant decay mode is therefore $\t \tbar$, if kinematically
allowed. An effective $\g\g$--coupling arises through an anomaly,
and is roughly comparable in size with
that to $\b \bbar$.
Techni-$\eta$ production at hadron colliders is therefore
predominantly through $\g \g$ fusion, as implemented in process 149.
In topcolor-assisted technicolor (discussed below), particles
like the techni-$\eta$ should not have a predominant coupling
to $\t$ quarks. In this sense, the process is considered obsolete.
(The following discussion borrows liberally from the introduction
to Ref.~\cite{Lan99a} with the author's permission.)
Modern technicolor models require
walking technicolor~\cite{Hol81} to prevent
large flavor-changing neutral currents
and the assistance of topcolor (TC2) interactions that are strong near
1~TeV~\cite{Nam88,Hil95,Lan95} to provide the large mass of the
top quark. Both additions to the basic technicolor
scenario~\cite{Wei79,Eic80} tend to require a large number $N_D$
of technifermion doublets to make the $\beta$-function of walking
technicolor small. They are needed in TC2 to generate
the hard masses of quarks and leptons, to induce the right mixing between
heavy and light quarks, and to break topcolor symmetry down to ordinary
colour. A large number of
techni-doublets implies a relatively low technihadron
mass scale \cite{Lan89,Eic96}, set by the
technipion decay constant $F_T \simeq F_\pi/\sqrt{N_D}$, where
$F_\pi = 246$ GeV.
The model adopted in {\Py} is the `Technicolor
Straw Man Model' (TCSM) \cite{Lan99a,Lan02a}.
The TCSM describes the phenomenology
of color-singlet vector and pseudoscalar technimesons and their
interactions with SM particles.
These technimesons are expected to be the lowest-lying
bound states of the lightest technifermion doublet, $(T_U, T_D)$,
with components that
transform under technicolor ${\bf SU(\Ntc)}$ as
fundamentals, but are QCD singlets; they have electric charges $Q_U$ and
$Q_D=Q_U-1$.
The vector technimesons form a spin-one isotriplet $\tro^{\pm,0}$ and an
isosinglet $\tom$. Since techni-isospin is likely to be a good approximate
symmetry, $\tro$ and $\tom$ should be approximately
mass-degenerate. The pseudoscalars,
or technipions, also comprise an isotriplet $\Pi_\mrm{tc}^{\pm,0}$
and an isosinglet
$\Pi_\mrm{tc}^{0 \prime}$. However, these are not mass eigenstates. In this
model, they are simple, two-state mixtures of the longitudinal
weak bosons $W_L^\pm$, $Z_L^0$ --- the true Goldstone bosons of dynamical
electroweak symmetry breaking in the limit that the
{$\bf SU(2) \otimes U(1)$} couplings $g,g'$ vanish --- and mass-eigenstate
pseudo-Goldstone technipions $\tpi^\pm, \tpiz$:
\be\label{eq:pistates}
\vert\Pi_\mrm{tc}\rangle = \sin\chi \ts \vert
W_L\rangle + \cos\chi \ts \vert\tpi\rangle\ts;\,
\vert\Pi_\mrm{tc}^{0 \prime} \rangle
= \cos\chipr \ts \vert\tpipr\rangle\ + \cdots,
\end{equation}
where $\sin\chi = F_T/F_\pi \ll 1$, $\chi'$ is another
mixing angle and the ellipsis refer to other technipions needed to
eliminate the TC anomaly from the $\Pi_\mrm{tc}^{0 \prime}$ chiral
current. These massive technipions are also expected to be approximately
degenerate.
The coupling of technipions to quarks and leptons
are induced mainly by extended technicolor (ETC)
interactions~\cite{Eic80}.
These couplings are proportional to fermion mass, except for
the case of the top quark, which has most of its mass generation through
TC2 interactions. The coupling to electroweak gauge boson pairs vanishes
at tree-level, and is assumed to be negligible. Thus
the ordinary mechanisms for producing Higgs-like bosons through
enhanced couplings to heavy fermions or heavy gauge bosons is absent for
technipions. In the following, we will concentrate on how
technipions decay once they are produced.
Besides coupling to fermions proportional
to mass (except for the case of top quarks where the coupling
strength should be much less than $m_\t$),
the $\tpipr$ can decay to gluon or photon pairs through
technifermion loops.
However,
there may be appreciable $\tpiz$--$\tpipr$ mixing \cite{Eic96}. If that
happens, the lightest neutral technipions are ideally-mixed $\bar T_U T_U$
and $\bar T_D T_D$ bound states. To simulate this effect, there are
separate factors $C_{\tpiz\to \g\g}$ and $C_{\tpipr\to \g\g}$ to weight
the $\tpi$ and $\tpipr$ partial widths for $\g\g$ decays.
The relevant technipion
decay modes are $\tpip \ra \t \bar \b, \c \bar \b, \u \bar \b$,
$\c \bar \s$, $\c \bar \d$ and $\tau^+ \nu_\tau$;
$\tpiz \ra \t \bar\t, \b \bar \b$, $\c\bar\c$, and $\tau^+\tau^-$; and
$\tpipr \ra \g\g$, $\t \bar\t, \b \bar \b$, $\c \bar \c$, and
$\tau^+\tau^-$. In the numerical evaluation of partial widths, the
running mass (see \ttt{PYMRUN}) is used, and all fermion pairs
are considered as final states.
The decay $\tpip\to \W^+\b\bar\b$ is also included,
with the final-state kinematics distributed according to phase
space (i.e.\ not weighted by the squared matrix element).
The $\tpi$ couplings to fermions can be weighted by parameters
$C_\c$, $C_\b$, $C_\t$ and $C_\tau$ depending on the heaviest
quark involved in the decay.
The technivector mesons have direct couplings to the technipion interaction
states.
In the limit of vanishing gauge couplings $g,g' = 0$,
the $\tro$ and $\tom$ coupling to technipions are:
\bea\label{eq:vt_decays}
\tro &\ra& \Pi_\mrm{tc} \Pi_\mrm{tc} = \cos^2 \chi\ts (\tpi\tpi) +
2\sin\chi\ts\cos\chi
\ts (\W_L\tpi) + \sin^2 \chi \ts (\W_L \W_L) \ts; \nn \\
\tom &\ra& \Pi_\mrm{tc} \Pi_\mrm{tc} \Pi_\mrm{tc} =
\cos^3 \chi \ts (\tpi\tpi\tpi) + \cdots \ts.
\eea
The $\tro\to\tpi\tpi$ decay amplitude, then, is given simply by
\be\label{eq:rhopipi}
\CM(\tro(q) \ra \pi_A(p_1) \pi_B(p_2)) = g_{\tro} \ts \CC_{AB}
\ts \epsilon(q)\cdot(p_1 - p_2) \ts,
\end{equation}
where the technirho coupling
$\atro \equiv g_{\tro}^2/4\pi = 2.91(3/\Ntc)$ is scaled na\"{\i}vely from
QCD ($\Ntc=4$ by default) and $\CC_{AB} = \cos^2\chi$ for $\tpi\tpi$,
$\sin\chi \cos\chi$ for $\tpi \W_L$, and $\sin^2\chi$ for $\W_L \W_L$.
While the technirho couples to $\W_L\W_L$, the coupling is suppressed.
Technivector production will be
addressed shortly; here, we concentrate on technivector decays.
Walking technicolor enhancements of technipion masses are assumed to close
off the channel $\tom \ra \tpi\tpi\tpi$ (which is not included) and
to kinematically suppress the channels $\tro \ra \tpi\tpi$ and the
isospin-violating $\tom \ra \tpi\tpi$ (which are allowed with appropriate
choices of mass parameters).
The rates for the isospin-violating decays $\tom \ra \pi^+_A \pi^-_B =
\W^+_L \W^-_L$, $\W^\pm_L \tpimp$, $\tpip \tpim$ are given by
$\Gamma(\tom \ra \pi^+_A \pi^-_B) = \vert\epsilon_{\rho\omega}\vert^2 \ts
\Gamma(\troz \ra \pi^+_A \pi^-_B)$
where $\epsilon_{\rho\omega}$ is the isospin-violating $\tro$-$\tom$ mixing.
Based on analogy with QCD, mixing of about $5\%$ is expected.
Additionally, this decay mode is dynamically suppressed, but it is
included as a possibility.
While a light technirho can
decay to $\W_L \tpi$ or $\W_L \W_L$ through TC dynamics,
a light techniomega decays mainly
through electroweak dynamics, $\tom \ra \gamma\tpiz$,
$Z^0\tpiz$, $W^\pm \tpimp$, etc., where $\Z$ and $\W$ may are transversely
polarized. Since
$\sin^2\chi \ll 1$, the electroweak decays of $\tro$ to the transverse gauge
bosons $\gamma,\W,\Z$ plus a technipion may be competitive with the
open-channel strong decays.
Note, the exact meaning of longitudinal or transverse polarizations only
makes sense at high energies, where the Goldstone equivalence theorem
can be applied.
At the moderate energies considered in the TCSM,
the decay products of the $\W$ and $\Z$ bosons are distributed according
to phase space, regardless of their designation as longitudinal
$\W_L/\Z_L$ or ordinary transverse gauge bosons.
To calculate the rates for transverse gauge boson decay,
an effective Lagrangian for technivector interactions
was constructed \cite{Lan99a}, exploiting gauge invariance,
chiral symmetry, and angular momentum and parity conservation.
As an example, the lowest-dimensional operator
mediating the decay $\tom(q) \ra \gamma(p_1) \tpiz(p_2))$
is $(e/M_V)\ts F_{\tro} \cdot \widetilde{F}_\gamma \ts
\tpiz$, where the mass parameter $M_V$ is expected to be
of order several 100~GeV. This leads to the
decay amplitude:
\be\label{eq:omgampi}
\CM(\tom(q) \ra \gamma(p_1) \tpiz(p_2)) = {e \cos\chi\over{M_V}}
\epsilon^{\mu\nu\lambda\rho} \epsilon_\mu(q) \epsilon^*_\nu(p_1) q_\lambda
p_{1\rho} \ts.
\end{equation}
Similar expressions exist for the other amplitudes involving different
technivectors and/or different gauge bosons \cite{Lan99a}, where
the couplings are derived in the valence technifermion
approximation \cite{Eic96,Lan99}.
In a similar fashion, decays to fermion-antifermion pairs are included.
These partial widths are typically small, but can have important
phenomenological consequences, such as narrow lepton-antilepton
resonances produced with electroweak strength.
Next, we address the issue of techniparticle production.
Final states containing
Standard Model particles and/or pseudo-Goldstone bosons (technipions)
can be produced at colliders
through two mechanisms: technirho and techniomega mixing
with gauge bosons through a vector-dominance mechanism,
and anomalies \cite{Lan02} involving
technifermions in loops.
Processes 191, 192 and 193 are based on $s$-channel production of
the respective resonance \cite{Eic96}
in the narrow width approximation.
All decay modes implemented can
be simulated separately or in combination, in the standard fashion.
These include pairs of fermions, of gauge bosons, of technipions,
and of mixtures of gauge bosons and technipions.
Processes 194, 195 and 361--377,
instead, include interference, a correct treatment
of kinematic thresholds and the anomaly contribution,
all of which can be important effects, but
also are limited to specific final states. Therefore, several processes
need to be simulated at once to determine the full effect of TC.
Process 194 is intended to more accurately represent the
mixing between the $\gamma^*$, $\Z^0$, $\rho_{\mrm{tc}}^0$ and
$\omega_{\mrm{tc}}^0$ particles in the Drell--Yan process \cite{Lan99}.
Process 195 is the analogous charged channel process including
$\W^{\pm}$ and $\rho_{\mrm{tc}}^{\pm}$ mixing. By default, the
final-state fermions are $\e^+ \e^-$ and $\e^{\pm} \nu_{\e}$,
respectively. These can be changed through the parameters
\ttt{KFPR(194,1)} and \ttt{KFPR(195,1)}, respectively (where the
\ttt{KFPR} value should represent a charged fermion).
Processes 361--368 describe the pair production of technipions and
gauge bosons through
$\rho_{\mrm{tc}}^0/\omega_{\mrm{tc}}^0$ resonances and anomaly contributions.
Processes 370--377 describe pair production through the
$\rho_{\mrm{tc}}^{\pm}$ resonance and anomalies.
It is important to note that processes \ttt{361, 362, 370, 371, 372}
include final states with
only longitudinally-polarized $\W$ and $\Z$ bosons,
whereas the others include final states with only transverse $\W$
and $\Z$ bosons. Again, {\bf all} processes must be simulated to
get the {\bf full} effect of the TC model under investigation.
All processes 361--377 are obtained by setting \ttt{MSEL = 50}.
The vector dominance mechanism is
implemented using the full
$\gamma$--$\Z^0$--$\tro$--$\tom$ propagator matrix, $\Delta_0(s)$, including
the effects of kinetic mixing. With the notation
$\CM^2_V = M^2_V - i \sqrt{s} \ts \Gamma_V(s)$ and $\Gamma_V(s)$ the
energy-dependent width for $V = \Z^0,\tro,\tom$, this
matrix is the inverse of
\begin{equation}
\label{eq:gzprop}
\Delta_0^{-1}(s) =\left(\ba{cccc}
s & 0 & s f_{\gamma\tro} & s f_{\gamma\tom} \\
0 & s - \CM^2_\Z & s f_{\Z\tro} & s f_{\Z\tom} \\
s f_{\gamma\tro} & s f_{\Z\tro} & s - \CM^2_{\troz} & 0 \\
s f_{\gamma\tom} & s f_{\Z\tom} & 0 & s - \CM^2_{\tom}
\ea\right) \ts.
\end{equation}
The parameters $f_{\gamma\tro} = \xi$, $f_{\gamma\tom} = \xi \ts (Q_U + Q_D)$,
$f_{\Z\tro} = \xi \ts \cot 2\thw$, and $f_{\Z\tom} = - \xi \ts
(Q_U + Q_D) \tan\thw$, and $\xi = \sqrt{\alpha/\atro}$ determine
the strength of the kinetic mixing, and are fixed by the
quantum numbers of the technifermions in the theory.
Because of the off-diagonal entries, the propagators resonate
at mass values shifted from the nominal $M_V$ values. Thus, while
users input the technihadron masses using {\ttt{PMAS}} values, these
will not represent exactly the resulting mass spectrum of
pair-produced particles.
Note that special care is taken in the
limit of very heavy technivectors to reproduce the canonical
$\gamma^*/\Z^*\to\tpip\tpim$ couplings.
In a similar fashion,
cross sections for charged final states require the $\W^\pm$--$\tropm$
matrix $\Delta_{\pm}$:
\be
\label{eq:wprop}
\Delta_{\pm}^{-1}(s) =\left(\ba{cc} s - \CM^2_\W & s f_{\W\tro} \\ s
f_{\W\tro} & s - \CM^2_{\tropm} \ea\right) \ts,
\end{equation}
where $f_{\W\tro} = \xi/(2\sin\thw)$.
By default, the TCSM Model has the parameters
$\Ntc$= 4,
$\sin\chi$ = $1\over 3$,
$Q_U$ = $4 \over 3$,
$Q_D = Q_U-1$ = $1\over 3$,
$C_\b=C_\c=C_\tau$= 1,
$C_\t$= $m_\b/m_\t$,
$C_{\tpi}$= $\tx{4\over{3}}$,
$C_{\tpiz\to \g\g}$=0, $C_{\tpipr\to \g\g}$=1,
$\vert\epsilon_{\rho\omega}\vert$ = 0.05,
$F_T = F_\pi \sin\chi$ = $82$ GeV,
$M_{\tropm}= M_{\troz} = M_{\tom}$ = $210$ GeV,
$M_{\tpipm} = M_{\tpiz} = M_{\tpipr}$ = $110$ GeV,
$M_{V} = M_{A}$ = $200$ GeV.
The techniparticle mass parameters are set through
the usual \ttt{PMAS} array.
Parameters regulating production and decay rates are stored in
the \ttt{RTCM} array in \ttt{PYTCSM}. This concludes the
discussion of the electroweak sector of the strawman model.
In the original TCSM outlined above, the existence of top-color interactions
only affected the coupling of technipions to top quarks, which
is a significant effect only for higher masses. In general, however,
TC2 requires some new and possibly light coloured particles.
In most TC2 models, the existence of a large $\t\bar\t$, but
not $\b\bar\b$, condensate and mass is due to $\suone \otimes \uone$ gauge
interactions which are strong near 1~TeV. The $\suone$ interaction is
$\t$--$\b$ symmetric while $\uone$ couplings are
$\t$--$\b$ asymmetric. There are
weaker $\sutwo \otimes \utwo$ gauge interactions in which light quarks (and
leptons) may \cite{Hil95}, or may not~\cite{Chi96}, participate. The two
${\bf U(1)}$'s must be
broken to weak hypercharge ${\bf U(1)_Y}$ at an energy somewhat
higher than 1~TeV
by electroweak-singlet condensates.
The full phenomenology of even such a simple model can be quite complicated,
and many (possibly unrealistic) simplifications are made to reduce
the number of free parameters \cite{Lan02a}. Nonetheless, it is useful to
have some benchmark to guide experimental searches.
The two TC2 ${\bf SU(3)}$'s can be broken to their diagonal $\suc$
subgroup by using technicolor and $\uone$ interactions, both strong near
1~TeV. This can be explicitly accomplished
\cite{Lan95} using two
electroweak doublets of technifermions, $T_1= (U_1,D_1)$ and $T_2 =
(U_2,D_2)$, which transform respectively as $(3,1,\Ntc)$ and $(1,3,\Ntc)$
under the two colour groups and technicolor. The desired pattern of symmetry
breaking occurs if ${\bf SU(\Ntc)}$ and $\uone$ interactions work together to
induce electroweak and $\suone \otimes \sutwo$ non-invariant condensates
$\langle \ts \bar U_{iL} U_{jR} \rangle$ and $\langle \ts
\bar D_{iL} D_{jR} \rangle$, $(i,j = 1,2)\ts$.
This minimal TC2 scenario leads to a rich spectrum of colour-nonsinglet
states readily accessible in hadron collisions. The lowest-lying ones
include eight `colorons', $V_8$, the massive gauge bosons of broken
topcolor ${\bf SU(3)}$;
four isosinglet $\troct$ formed from $\bar T_i T_j$ and the
isosinglet pseudo-Goldstone technipions formed from $\bar T_2 T_2$.
In this treatment, the isovector technipions are ignored, because
they must be pair produced in $\troct$ decays,
and such decays are assumed to be
kinematically suppressed.
The colorons are new fundamental particles with couplings to quarks.
In standard TC2~\cite{Hil95}, top and bottom quarks couple to $\suone$
and the four light quarks to $\sutwo$. Because the
$\suone$ interaction is strong and acts exclusively on the third
generation, the residual $V_8$ coupling can be enhanced for $\t$
and $\b$ quarks. The coupling $g_a=g_c\cot\theta_3$ for $\t$ and $\b$
and $g_a=-g_c\tan\theta_3$ for $\u,\d,\c,\s$,
where $g_c$ is the QCD coupling and $\cot\theta_3$ is related to the
original $g_1$ and $g_2$ couplings.
In flavor-universal TC2~\cite{Chi96} all quarks couple to
$\suone$, not $\sutwo$, so that colorons couple equally and strongly to all
flavors: $g_a = g_c \cot\theta_3$.
Assuming that techni-isospin is not badly broken by ETC interactions,
the $\troct$ are isosinglets
labeled by the
technifermion content and colour index $A$:
$\rho_{11}^{A}, \rho_{22}^{A}, \rho_{12}^{A}, \rho_{12'}^{A}$.
The first two of these states, $\troctaa$ and $\troctbb$, mix with $V_8$ and
$g$. The topcolor-breaking condensate, $\langle \bar T_{1L} T_{2R} \rangle
\neq 0$, causes them to also mix with $\troctab$ and $\troctabp$.
Technifermion condensation also leads to a number of (pseudo)Goldstone boson
technipions. The lightest technipions are expected to be
the isosinglet $\suc$ octet and singlet $\bar T_2
T_2$ states $\pi_{22}^{A}$ and $\pi_{22}^0$.
These technipions can decay into either
fermion-antifermion pairs or two gluons; presently, they are
assumed to decay only into gluons. As noted, walking
technicolor enhancement of technipion
masses very likely close off the $\troct \ra
\tpi\tpi$ channels. Then the $\troct$ decay into $\q\bar\q$ and $\g\g$.
The rate for the former are proportional to the amount
of kinetic mixing, set by
$\xi_\g = {g_c / {g_{\tro}}}$.
Additionally, the $\troctbb$ decays to $\g\pi_{22}^{0,A}$.
The $V_8$ colorons are expected to be considerably heavier than the $\troct$,
with mass in the range 0.5--1~TeV. In both the standard and
flavor-universal models, colorons couple strongly to $\bar T_1 T_1$, but with
only strength $g_c$ to $\bar T_2 T_2$. Since relatively light technipions are
$\bar T_2 T_2$ states, it is estimated that
$\Gamma(V_8 \ra \tpi\tpi) = \CO(\alpha_c)$
and $\Gamma(V_8 \ra g\tpi) = \CO(\alpha_c^2)$. Therefore, these decay
modes are ignored, so that
the $V_8$ decay rate is the
sum over open channels of
\be\label{eq:vdecay}
\Gamma(V_8 \ra \q_a \bar \q_a) = {\alpha_a \over{6}} \ts
\left(1 + {2m_a^2 \over{s}} \right) \ts
\left(s - 4m_a^2\right)^{1\over{2}} \ts,
\end{equation}
where $\alpha_a = g^2_a/4\pi$.
The phenomenological effect of this techniparticle structure is
to modify the gluon propagator in ordinary QCD processes, because
of mixing
between the gluon, $V_8$ and the $\troct$'s.
The $\g$--$V_8$--$\troctaa$--$\troctbb$--$\troctab$--$\troctabp$ propagator
is the inverse of the symmetric matrix
\be\label{eq:invprop}
D^{-1}(s) = \left(\ba{cccccc}
s & 0 & s \ts \xi_\g & s \ts \xi_\g & 0 & 0 \\ \\
0 & s - \CM^2_{V_8} & s \ts \xi_{\rho_{11}} & s \ts \xi_{\rho_{22}} & s
\ts \xi_{\rho_{12}} & s \ts \xi_{\rho_{12'}}\\ \\
s \ts \xi_\g & s \ts \xi_{\rho_{11}} & s - \CM^2_{11} & - M^2_{11,22} &
-M^2_{11,12} & -M^2_{11,12'}\\ \\
s \ts \xi_\g & s \ts \xi_{\rho_{22}} & -M^2_{11,22} & s - \CM^2_{22} &
-M^2_{22,12} & -M^2_{22,12'}\\ \\
0 & s \ts \xi_{\rho_{12}} & -M^2_{11,12} & -M^2_{22,12} &
s - \CM^2_{12} & -M^2_{12,12'}\\ \\
0 & s \ts \xi_{\rho_{12'}} & -M^2_{11,12'} & -M^2_{22,12'} &
-M^2_{12,12'} & s - \CM^2_{12'} \\
\ea\right) \ts.
\end{equation}
Here, $\CM^2_V = M^2_V - i \sqrt{s} \ts \Gamma_V(s)$ uses the
energy-dependent widths of the octet vector bosons, and
the $\xi_{\rho_{ij}}$ are proportional to $\xi_\g$ and
elements of matrices that describe the pattern of technifermion
condensation. The mixing terms $M^2_{ij,kl}$, induced by
$\bar T_1 T_2$ condensation are assumed to be real.
This extension of the TCSM is still under development, and
any results should be carefully scrutinized.
The main effects are indirect, in that they modify
the underlying two-parton QCD processes much like compositeness terms,
except that a resonant structure is visible.
Similar to compositeness, the effects of these colored technihadrons
are simulated by setting \ttt{ITCM(5) = 5} for processes 381--388.
By default, these processes are equivalent to the 11, 12, 13, 28, 53,
68, 81 and 82 ones, respectively. The last two are specific for
heavy-flavour production, while the first six could be used to describe
standard or non-standard high-$\pT$ jet production. These six are
simulated by \ttt{MSEL = 51}.
The parameter dependence of the `model' is encoded
in $\tan\theta_3$ (\ttt{RTCM(21)}) and a mass parameter
$M_8$ (\ttt{RTCM(27)}), which determines the decay width
$\rho_{22}\to \g\pi_{22}$ analogously to $M_V$ for $\tom\to\gamma\tpi$.
For \ttt{ITCM(2)} equal to 0 (1), the
standard (flavor universal) TC2 couplings are used.
The mass parameters are set by the \ttt{PMAS} array using the
codes: $V_8$ (3100021), $\pi_{22}^{1}$ (3100111),
$\pi_{22}^{8}$ (3200111), $\rho_{11}$ (3100113),
$\rho_{12}$ (3200113), $\rho_{21}$ (3300113), and $\rho_{22}$ (3400113).
The mixing parameters $M_{ij,kl}$ take on the (arbitrary) values
$M_{11,22}=100$ GeV, $M_{11,12}=M_{11,21}=M_{22,12}=150$ GeV,
$M_{22,21}=75$ GeV and $M_{12,21}=50$ GeV, while the
kinetic mixing terms $\xi_{\rho_{ij}}$ are calculated assuming
the technicolor condensates are fully mixed, i.e.\
$\langle T_i\bar T_j \rangle \propto {1 / \sqrt{2}}$.
\subsubsection{Extra Dimensions}
\label{sss:extradimclass}
{\ISUB} =
\begin{tabular}[t]{rl}
391 & $\f \fbar \to \G^*$ \\
392 & $\g \g \to \G^*$ \\
393 & $\q \qbar \to \g \G^*$ \\
394 & $\q \g \to \q \G^*$ \\
395 & $\g \g \to \g \G^*$ \\
\end{tabular}
In recent years, the area of observable consequences of extra dimensions
has attracted a strong interest. The field is still in rapid development,
so there still does not exist a `standard phenomenology'. The topic is
also fairly new in {\Py}, and currently only a first scenario is
available.
The $\G^*$, introduced as new particle code 5000039, is intended
to represent the lowest excited graviton state in a Randall-Sundrum
scenario \cite{Ran99} of extra dimensions. The lowest-order production
processes by fermion or gluon fusion are found in 391 and 392. The
further processes 393--395 are intended for the high-$\pT$ tail in hadron
colliders. As usual, it would be double-counting to have both sets of
processes switched on at the same time. Processes 391 and 392, with
initial-state showers switched on, are appropriate for the full
cross section at all $\pT$ values, and gives a reasonable description
also of the high-$\pT$ tail. Processes 393--395 could be useful e.g.
for the study of invisible decays of the $\G^*$, where a large $\pT$
imbalance would be required. It also serves to test/confirm the shower
expectations of different $\pT$ spectra for different production
processes \cite{Bij01}.
Decay channels of the $\G^*$ to $\f \fbar$, $\g\g$, $\gamma \gamma$,
$\Z^0 \Z^0$ and $\W^+ \W^-$ contribute to the total width. The correct
angular distributions are included for decays to a fermion, $\g$ or
$\gamma$ pair in the lowest-order processes, whereas other decays
currently are taken to be isotropic.
The $\G^*$ mass is to be considered a free parameter. The other degree
of freedom in this scenario is a dimensionless coupling; see
\ttt{PARP(50)}.
\subsection{Supersymmetry}
\label{ss:susyclass}
\ttt{MSEL} = 39--45 \\
{\ISUB} = 201--296 (see tables at the beginning of this chapter)
{\Py} includes the possibility of simulating a large variety of production
and decay processes in the Minimal Supersymmetric extension
of the Standard Model (MSSM). The simulation is based
on an effective Lagrangian of softly-broken SUSY with parameters
defined at the weak scale, typically between $m_{\Z}$ and 1
TeV. The relevant parameters should either be supplied directly by the user or
they should be read in from a SUSY Les Houches Accord (SLHA) spectrum
file \cite{Ska03}
(see below). Some other possibilities for obtaining the SUSY parameters
also exist in the code, as described below, but these
are only intended for backwards compatibility and debugging purposes.
\subsubsection{General Introduction}
In any ($N=1$)
supersymmetric version of the SM there exists a partner to each SM
state with the same gauge quantum numbers but whose spin differs by one half
unit. Additionally, the dual requirements of generating masses for
up- and down-type fermions while preserving SUSY and gauge
invariance, require that the SM Higgs sector be enlarged to two scalar
doublets, with corresponding spin-partners.
After Electroweak symmetry breaking (EWSB), the bosonic Higgs sector contains
a quintet of physical states: two CP-even scalars, $\hrm^0$ and $\H^0$, one
CP-odd pseudoscalar, $\A^0$, and a pair of charged scalar Higgs bosons,
$\H^\pm$ (naturally, this classification is only correct when CP violation
is absent in the Higgs sector. Non-trivial phases between certain
soft-breaking parameters will induce mixing between the CP eigenstates).
The fermionic Higgs (called `Higgsino') sector is constituted by
the superpartners of these fields, but these are not normally exact mass
eigenstates, so we temporarily postpone the discussion of them.
In the gauge sector, the spin-$1/2$ partners of the $\bf U(1)_Y$ and $\bf
SU(2)_L$ gauge bosons (called `gauginos') are the Bino, $\Bino$, the
neutral Wino, $\w3$, and the charged Winos, $\Wino_1$ and $\Wino_2$, while
the partner of the gluon is the gluino, $\gluino$. After EWSB, the $\Bino$
and $\w3$ mix with the neutral Higgsinos, $\higgsino_1, \higgsino_2$, to form
four neutral Majorana fermion mass-eigenstates, the neutralinos,
$\zinog_{1-4}$. In addition, the charged Higgsinos, $\higgsino^\pm$, mix with
the charged Winos, $\Wino_1$ and $\Wino_2$, resulting in two charged Dirac
fermion mass eigenstates, the charginos, $\winog_{1,2}$. Note that the
$\tilde{\gamma}$ and $\tilde{\Z}$, which sometimes occur in the literature,
are linear combinations of the $\Bino$ and $\w3$, by exact analogy with the
mixing giving the $\gamma$ and $\Z^0$, but these are not normally mass
eigenstates after EWSB, due to the enlarged mixing caused by the presence of
the Higgsinos.
The spin-0 partners of the SM fermions (so-called `scalar fermions', or
`sfermions') are the squarks $\squark$, sleptons $\slepton$, and sneutrinos
$\sneutrino$. Each fermion (except possibly the neutrinos) has two scalar
partners, one associated with each of its chirality states. These are named
left-handed and right-handed sfermions, respectively. Due to their scalar
nature, it is of course impossible for these particles to possess any
intrinsic `handedness' themselves, but they inherit their couplings to the
gauge sector from their SM partners, so that e.g.\ a $\tilde \d_R$ does not
couple to {$\bf SU(2)_L$} while a $\tilde \d_L$ does.
Generically, the \ttt{KF} code numbering scheme used in {\Py} reflects the
relationship between particle and sparticle, so that e.g.~for sfermions, the
left-handed (right-handed) superpartners have codes 1000000 (2000000) plus
the code of the corresponding SM fermion. A complete list of the particle
partners and their {\KF} codes is given in Table~\ref{t:codenine}. Note that,
antiparticles of scalar particles are denoted by $^*$, i.e.\
$\st^*$. A gravitino is also included with \ttt{KF=1000039}. The gravitino
is only relevant in {\Py} when simulating models of gauge-mediated SUSY
breaking,
where the gravitino becomes the lightest superpartner. In practice,
the gravitino simulated here is the spin-$1\over 2$ Goldstino components
of the spin-$3\over 2$ gravitino.
The MSSM Lagrangian contains interactions between particles
and sparticles, with couplings fixed by SUSY. There are also a number of
soft SUSY-breaking mass parameters. `Soft' here means
that they break the
mass degeneracy between SM particles and their SUSY partners
without reintroducing quadratic divergences in the theory or destroying
its gauge invariance.
In the MSSM, the soft SUSY-breaking parameters are extra mass terms for
gauginos and sfermions and trilinear scalar couplings. Further soft terms may
arise, for instance in models with broken $R$-parity, but we here restrict
our attention to the minimal case (for RPV in \Py\ see section
\ref{sss:rparityviol}).
The exact number of independent parameters
depends on the detailed mechanism of SUSY breaking.
The general MSSM model in {\Py} assumes only a few relations
between these parameters which seem theoretically difficult to
avoid. Thus, the first two generations of sfermions with
otherwise similar quantum numbers, e.g.\ $\tilde \d_L$ and $\tilde \s_L$,
have the same masses.
Despite such simplifications, there are a fairly large number of
parameters that appear in the SUSY Lagrangian and determine
the physical masses and interactions with Standard Model particles,
though far less than the $> 100$ which are allowed in all generality.
The Lagrangian (and, hence, Feynman rules) follows the conventions
set down by Kane and Haber in their Physics Report article
\cite{Hab85} and the papers of Gunion and Haber \cite{Gun86a}.
Once the parameters of the softly-broken SUSY Lagrangian are
specified, the interactions are fixed, and the sparticle masses can
be calculated. Note that, when using SUSY Les Houches Accord input,
\Py\ automatically translates between the SLHA conventions and the
above, with no action required on the part of the user.
\subsubsection{Extended Higgs Sector}
{\Py} already simulates a Two Higgs Doublet Model (2HDM) obeying tree-level
relations fixed by two parameters, which can be conveniently taken as the
ratio of doublet vacuum expectation values $\tan\beta$, and the
pseudoscalar mass $M_{\A}$ (as noted earlier, for the non-SUSY
implementation of a 2HDM, the input parameters are $M_\hrm$ and $\tan\beta$).
The Higgs particles are considered Standard Model fields, since a
2HDM is a straightforward extension of the Standard Model. The MSSM Higgs
sector is more complicated than that described above in
section \ref{ss:Hclass}, and includes important radiative corrections to
the tree-level relations. The CP-even Higgs mixing angle $\alpha$ is
shifted as well as the full Higgs mass spectrum. The properties of the
radiatively-corrected Higgs sector in {\Py} are derived in the effective
potential approach \cite{Car95}. The effective potential contains an
all-orders resummation of the most important radiative corrections, but
makes approximations to the virtuality of internal propagators. This is to
be contrasted with the diagrammatic technique, which performs a fixed-order
calculation without approximating propagators. In practice, both techniques
can be systematically corrected for their respective approximations, so that
there is good agreement between their predictions, though sometimes the
agreement occurs for slightly different values of SUSY-breaking
parameters. The calculation of the Higgs spectrum in {\Py}
is based on the {\ttt{FORTRAN}} code {\ttt{SubHpole}} \cite{Car95},
which is also used in \ttt{HDecay} \cite{Djo97}, except that certain
corrections that are particularly important at large values of $\tan\beta$
are included rigorously in {\Py}.
There are several notable properties of the MSSM Higgs sector. As long as
the soft SUSY-breaking parameters are less than about 1.5 TeV, a number
which represents a fair, albeit subjective, limit for where the required
degree of fine-tuning of MSSM parameters becomes unacceptably large, there is
an upper bound of about 135 GeV on the mass of the CP-even Higgs boson most
like the Standard Model one, i.e.\ the one with the largest couplings to the
$\W$ and $\Z$ bosons, be it the $\hrm$ or $\H$. If it is $\hrm$ that is the
SM-like Higgs boson, then $\H$ can be significantly heavier. On the other
hand, if $\H$ is the SM-like Higgs boson, then $\hrm$ must be even lighter.
If all SUSY particles are heavy, but $M_{\A}$ is small, then the
low-energy theory would look like a two-Higgs-doublet model.
For sufficiently large $M_{\A}$, the heavy Higgs doublet decouples,
and the effective low-energy theory has only one light Higgs doublet
with SM-like couplings to gauge bosons and fermions.
The Standard Model fermion masses are not fixed by SUSY,
but their Yukawa couplings become
a function of $\tan \beta$.
For the up- and down-quark and leptons,
$m_u = h_u v \sin \beta$,
$m_d = h_d v \cos \beta$, and
$m_\ell = h_\ell v \cos \beta$,
where $h_{f=u,d,\ell}$
is the corresponding Yukawa coupling and $v \approx 246$ GeV is the
order parameter of Electroweak symmetry breaking.
At large $\tan\beta$, significant corrections can occur
to these relations. These are included for the $\b$ quark,
which appears to have the most sensitivity to them, and
the $\t$ quark in the subroutine
\ttt{PYPOLE}, based on an updated version of \ttt{SubHpole},
which also includes some bug fixes, so that
it is generally better behaved.
The array values \ttt{RMSS(40)} and \ttt{RMSS(41)} are used for
temporary storage of the corrections $\Delta m_{\t}$ and $\Delta m_{\b}$.
The input parameters that determine the MSSM Higgs sector
in {\Py} are \ttt{RMSS(5)} ($\tan\beta$), \ttt{RMSS(19)} ($M_{\A}$),
\ttt{RMSS(10-12)} (the third generation squark mass parameters),
\ttt{RMSS(15-16)} (the third generation squark trilinear
couplings), and \ttt{RMSS(4)} (the Higgsino mass $\mu$).
Additionally, the large $\tan\beta$ corrections related
to the $\b$ Yukawa coupling depend on \ttt{RMSS(3)}
(the gluino mass).
Of course, these calculations also
depend on SM parameters ($m_{\t}, m_{\Z}, \alphas,$ etc.). Any
modifications to these quantities from virtual MSSM effects are not
taken into account. In principle, the sparticle masses also acquire
loop corrections that depend on all MSSM masses.
If \ttt{IMSS(4) = 0}, an approximate version of the effective potential
calculation can be used. It is not as accurate as that available for
\ttt{IMSS(4) = 1}, but it useful for demonstrating the effects of higher
orders. Alternatively, for \ttt{IMSS(4) = 2}, the physical Higgs masses
are set by their \ttt{PMAS} values while the CP-even Higgs boson mixing
angle $\alpha$ is set by \ttt{RMSS(18)}. These values and $\tan\beta$
(\ttt{RMSS(5)}) are enough to determine the couplings, provided that
the same tree-level relations are used.
See section \ref{sss:models} for a description how to use the
loop-improved RGE's of \tsc{Isasusy} to determine the SUSY
mass and mixing spectrum (including also loop
corrections to the Higgs mass spectrum and couplings) with \Py.
Finally, a run-time interface to \tsc{FeynHiggs} \cite{Hei99},
for the diagrammatic calculation of the $\hrm^0$, $\H^0$, $\A^0$, and $\H^+$
masses and the mixing angle $\alpha$ in the MSSM, has been introduced,
available through the option \ttt{IMSS(4) = 3}. For the time being, it can
be invoked either when using an SLHA SUSY spectrum, i.e.\ for
\ttt{IMSS(1) = 11}, or when using the run-time interface to \tsc{Isasusy},
i.e.\ for \ttt{IMSS(1) = 12} or 13. The interface calls three
\ttt{FeynHiggs} routines, in the following order:
\begin{Itemize}
\item \ttt{FHSETFLAGS(IERR,4,0,0,2,0,2,1,1) :} these are the `default'
settings recommended for \tsc{FeynHiggs} \cite{Hei99}.
\item \ttt{FHSETPARA :} to set the MSSM parameters.
\item \ttt{FHHIGGSCORR :} to get the corrected Higgs parameters.
\end{Itemize}
Note that, for {\Py} to compile properly without the \tsc{FeynHiggs}
library, three dummy routines have been added to the {\Py} source code,
corresponding to the three listed above. To obtain proper linking with
\tsc{FeynHiggs}, these dummy routines should first be removed/renamed
and the {\Py} source recompiled without them. The interface has been
tested to work with \tsc{FeynHiggs}-2.2.8.
Differences in the \tsc{FeynHiggs} and \ttt{SubHpole} predictions
represent, to some degree, the theoretical uncertainty in the
MSSM Higgs sector.
\subsubsection{Superpartners of Gauge and Higgs Bosons}
The chargino and neutralino masses and
their mixing angles (that is, their gaugino and Higgsino composition)
are determined by the SM gauge boson masses ($M_\W$ and $M_\Z$),
$\tan \beta$,
two soft SUSY-breaking
parameters (the ${\bf SU(2)_L}$ gaugino mass $M_2$ and the ${\bf U(1)_Y}$
gaugino mass $M_1$), together with the Higgsino mass parameter $\mu$,
all evaluated at the electroweak scale $\sim M_\Z$. {\Py} assumes
that the input parameters are evaluated at the `correct' scale.
Obviously, more care is needed to set precise experimental limits or
to make a connection to higher-order calculations.
Explicit solutions of the chargino and neutralino masses and
their mixing angles (which appear in Feynman rules)
are found by diagonalizing the $2\times 2$ chargino
$\mathbf{M_C}$ and
$4\times 4$ neutralino $\mathbf{M_N}$ mass matrices:
\begin{eqnarray}
\label{eqn:inomatrices}
&\mathbf{M_{C}} = \left( \begin{array}{cc}
M_2 & \sqrt{2}M_\W s\beta \\
\sqrt{2}M_\W c\beta & \mu \\ \end{array} \right);
\mathbf{M_{N}} = \left( \begin{array}{cc}
\mathbf{M}_i & \mathbf{Z} \\
\mathbf{Z^T} & \mathbf{M}_\mu \\ \end{array} \right) & \\
& \nonumber \\
&\mathbf{M}_i = \left( \begin{array}{cc}
M_1 & 0 \\
0 & M_2 \\ \end{array} \right);
\mathbf{M}_\mu = \left( \begin{array}{cc}
0 & -\mu \\
-\mu & 0 \\ \end{array} \right);
\mathbf{Z} = \left( \begin{array}{cc}
-M_\Z c\beta s_W & M_\Z s\beta s_W \\
M_\Z c\beta c_W & -M_\Z s\beta c_W \\ \end{array} \right) & \nonumber
\end{eqnarray}
$\mathbf{M_C}$ is written in the $(\Wino^+,\higgsino^+)$ basis,
$\mathbf{M_N}$ in the $(\Bino,\Wino^3,\higgsino_1,\higgsino_2)$ basis,
with the notation $s\beta=\sin\beta,c\beta=\cos\beta,s_W=\sin\theta_W$
and $c_W=\cos\theta_W$.
Different sign conventions and bases sometimes appear in
the literature. In particular, {\Py} agrees with the
{\sc Isasusy} \cite{Bae93} convention for $\mu$, but uses a different
basis of fields and has different-looking mixing matrices.
In general, the soft SUSY-breaking parameters can be
complex valued, resulting in CP violation in some sector of the theory, but
more directly expanding the possible masses and mixings of sparticles.
Presently, the consequences of arbitrary phases are
only considered in the chargino and neutralino sector,
though it is well known that they can have a significant
impact on the Higgs sector. A generalization of the
Higgs sector is among the plans for the future development
of the program. The chargino and neutralino
input parameters are \ttt{RMSS(5)} ($\tan\beta$),
\ttt{RMSS(1)} (the modulus of
$M_1$) and \ttt{RMSS(30)} (the phase of $M_1$),
\ttt{RMSS(2)} and \ttt{RMSS(31)} (the modulus and
phase of $M_2$), and \ttt{RMSS(4)} and \ttt{RMSS(33)}
(the modulus and phase of $\mu$). To simulate the
case of real parameters (which is CP-conserving),
the phases are zeroed by default. In addition,
the moduli parameters can be signed, to make a simpler
connection to the CP-conserving case. (For example,
\ttt{RMSS(5) = -100.0} and \ttt{RMSS(30) = 0.0} represents
$\mu=-100$ GeV.)
The expressions for the production cross sections
and decay widths
of neutralino and chargino pairs contain the phase dependence,
but ignore possible effects of the phases in the sfermion
masses appearing in propagators.
The production cross sections have been updated to include
the dependence on beam polarization through the
parameters \ttt{PARJ(131,132)} (see Sect.~\ref{ss:polarization}).
There are several approximations made for three-body decays.
The numerical expressions for three-body decay widths
ignore the effects of finite
fermion masses in the matrix element, but include them
in the phase space. No
three-body decays $\chi_i^0\to \t\bar \t\chi^0_j$ are simulated,
nor $\chi^+_i (\chi^0_i) \to \t \bar \b \chi^0_j (\chi^-_j)$.
Finally, the effects of mixing between the third generation interaction
and mass eigenstates for sfermions is ignored, except that the
physical sfermion masses are used.
The kinematic distributions of the decay products are spin-averaged,
but include the correct matrix-element weighting.
Note that for the $R$-parity-violating decays (see below), both
sfermion mixing effects and masses of $\b$, $\t$, and $\tau$ are fully
included.
In some corners of SUSY parameter space, special decay modes must be
implemented to capture important phenomenology. Three different
cases are distinguished here: (1) In the most common
models of Supergravity-mediated SUSY breaking, small values of
$M_1$, $M_2$, $\mu$ and $\tan\beta$ can lead to a neutralino spectrum with
small mass splittings. For this case, the radiative decays
$\chi^0_i \to \chi^0_j \gamma$ can be relevant, which are the SUSY analog
of $\hrm \to \gamma\gamma$ with two particles switched to superpartners
to yield $\widetilde\hrm \to \widetilde\gamma \gamma$, for example.
These decays are calculated approximately for all neutralinos
when $\tan\beta\le 2$, or the decay $\chi^0_2\to \chi^0_1\gamma$
can be forced using {\ttt{IMSS(10) = 1}}; (2) In models of gauge-mediated
SUSY breaking (GMSB), the gravitino $\widetilde\G$
is light and phenomenologically
relevant at colliders. For {\ttt{IMSS(11) = 1}}, the two-body
decays of sparticle to particle plus gravitino are allowed. The most
relevant of these decay modes are likely $\chi_1^0 \to \V\widetilde\G$,
with $\V=\gamma,\Z$ or a Higgs boson; (3) In models of anomaly-mediated
SUSY breaking (AMSB), the wino mass parameter $M_2$ is much smaller
than $M_1$ or $\mu$. As a result, the lightest chargino and neutralino
are almost degenerate in mass. At tree level, it can be shown
analytically that the chargino should be
heavier than the neutralino, but this is hard to achieve numerically.
Furthermore, for this case, radiative corrections are important in
increasing the mass splitting more. Currently,
if ever the neutralino is heavier than the chargino
when solving the
eigenvalue problem numerically, the chargino mass is set to the
neutralino mass plus 2 times the charged pion mass, thus allowing the
decay $\winog_1\to\pi^\pm\zinog_1$.
Since the ${\bf SU(3)_C}$ symmetry of the SM is not broken, the
gluinos have masses determined by the ${\bf SU(3)_C}$ gaugino mass
parameter $M_3$, input through the parameter \ttt{RMSS(3)}.
The physical gluino mass is shifted from the value
of the gluino mass parameter $M_3$ because of radiative corrections.
As a result, there is an indirect dependence on the squark masses.
Nonetheless, it is sometimes convenient to input the physical
gluino mass, assuming that there is some choice of $M_3$ which
would be shifted to this value. This can be accomplished through
the input parameter \ttt{IMSS(3)}.
A phase for the gluino mass can be set using \ttt{RMSS(32)},
and this can influence the gluino decay width (but no effect
is included in the $\gluino+\ino$ production).
Three-body decays of
the gluino to $\t \bar \t$ and $\b \bar \b$ and $\t \bar \b$
plus the appropriate neutralino or chargino are allowed and
include the full effects of sfermion mixing. However, they
do not include the effects of phases arising from complex
neutralino or chargino parameters.
\subsubsection{Superpartners of Standard Model Fermions}
The mass eigenstates of squarks and sleptons are, in principle,
mixtures of their left- and right-handed components, given by:
\begin{eqnarray}
M_{\tilde{f}_L}^2
= m_{\bf 2}^2 + m^2_f + D_{\tilde{f}_L} & &
M_{\tilde{f}_R}^2
= m_{\bf 1}^2 + m^2_f + D_{\tilde{f}_R}
\label{eq:LRmasseigen}
\end{eqnarray}
where $m_{\bf 2}$ are soft SUSY-breaking parameters
for superpartners of ${\bf SU(2)_L}$ doublets, and $m_{\bf 1}$
are parameters for singlets.
The $D$-terms
associated with Electroweak symmetry breaking are
$D_{\tilde{f}_L}= M_Z^2 \cos (2 \beta) (T_{3_{f}} -
Q_f \sin^2\theta_W)$ and
$D_{\tilde{f}_R}= M_Z^2 \cos( 2 \beta) Q_f\sin^2\theta_W$,
where $T_{3_f}$ is the weak isospin eigenvalue ($=\pm 1/2$) of the fermion and
$Q_f$ is the electric charge. Taking the $D$-terms into account, one easily
sees that the masses of sfermions in ${\bf SU(2)_L}$ doublets are related by
a sum rule: $M_{\tilde{f}_L,T_3=1/2}^2 - M_{\tilde{f}_L,T_3=-1/2}^2 =
M_Z^2\cos( 2 \beta)$.
In many high-energy models,
the soft SUSY-breaking sfermion mass parameters are
taken to be equal at the high-energy scale, but, in principle,
they can be different for
each generation or even within a generation.
However, the sfermion flavor dependence can have important effects on
low-energy observables, and it is often strongly constrained.
The suppression of flavor changing neutral currents (FCNC's), such as
$K_L\to\pi^\circ\nu\bar\nu$, requires that either $(i)$ the squark soft
SUSY-breaking mass matrix is diagonal and degenerate,
or $(ii)$ the masses of the
first- and second-generation sfermions are very large. Thus we make the
data-motivated simplification of setting $M_{\su_{L}}=M_{\sch_{L}}$,
$M_{\sd_{L}}=M_{\sst_{L}}$, $M_{\su_{R}}=M_{\sch_{R}}$,
$M_{\sd_{R}}=M_{\sst_{R}}$. \vspace{1mm}
The left--right
sfermion mixing is determined by the product of
soft SUSY-breaking parameters and the mass of
the corresponding fermion.
Unless the soft SUSY-breaking parameters for the first two
generations are orders of magnitude greater than for
the third generation, the mixing in
the first two generations can be neglected. This simplifying assumption is
also made in {\Py}: the sfermions $\squark_{L,R}$, with
$\squark = \tilde \u,\tilde \d,\tilde \c,\tilde \s$,
and $\slepton_{L,R},\sneutrino_\ell$,
with $\ell = \e, \mu$, are the real mass eigenstates with
masses $m_{\squark_{L,R}}$ and
$m_{\slepton_{L,R}}, m_{\sneutrino_{\ell}},$ respectively.
For the third generation sfermions, due to weaker experimental constraints,
the left--right mixing can be nontrivial.
The tree-level
mass matrix for the top squarks (stops) in the ($\stop_L, \stop_R$)
basis is given by
\begin{equation}
M^2_{\stop} = \left( \begin{array}{cc}
m_{Q_3}^2 + m_\t^2 + D_{\stop_L} & m_\t ( A_\t - \mu/ \tan \beta) \\
m_\t ( A_\t - \mu/ \tan \beta) & m_{U_3}^2 + m_\t^2 + D_{\stop_R} \\
\end{array} \right),
\label{stop_matrix}
\end{equation}
where $A_\t$ is a trilinear coupling.
Different sign conventions for $A_\t$
occur in the literature; {\Py} and \tsc{Isasusy} use
opposite signs. Unless there is
a cancellation between $A_\t$ and $\mu/\tan\beta$, left--right mixing
occurs for the stop squarks because of the large top quark mass.
The stop mass eigenstates are then given by
\begin{eqnarray}
\stop_1 & = &
\cos \theta_{\stop} \;\;\stop_L + \sin \theta_{\stop} \;\;
\stop_R
\nonumber \\
\stop_2 &= &
- \sin \theta_{\stop} \;\;\stop_L + \cos \theta_{\stop} \;\;
\stop_R,
\end{eqnarray}
where the masses and mixing angle $\theta_{\stop}$ are fixed by
diagonalizing the squared-mass matrix Eq.~(\ref{stop_matrix}).
Note that different conventions exist also for the mixing angle
$\theta_{\stop}$, and that {\Py} here agrees with \tsc{Isasusy}.
When translating Feynman rules from the (L,R) to (1,2) basis,
we use:
\begin{eqnarray}
\stop_L & = &
\cos \theta_{\stop} \;\;\stop_1 - \sin \theta_{\stop} \;\;
\stop_2
\nonumber \\
\stop_R &= &
\sin \theta_{\stop} \;\;\stop_1 + \cos \theta_{\stop} \;\;
\stop_2.
\end{eqnarray}
Because of the large mixing, the lightest stop $\stop_1$ can be
one of the lightest sparticles.
For the sbottom, an
analogous formula for the mass matrix holds with $m_{U_3}\to m_{D_3}$,
$A_\t\to A_\b$, $D_{\stop_{L,R}}\to D_{\sbottom_{L,R}}$, $m_\t\to m_\b$,
and $\tan\beta\to 1/\tan\beta$. For the stau, the substitutions
$m_{Q_3}\to m_{L_3}$,
$m_{U_3}\to m_{E_3}$, $A_\t\to A_\tau$, $D_{\stop_{L,R}}\to
D_{\stau_{L,R}}$,
$m_t\to m_\tau$ and $\tan\beta\to$ 1/$\tan\beta$ are appropriate.
The parameters
$A_\t$, $A_\b$, and $A_\tau$ can be independent,
or they might be related by some underlying principle.
When $m_\b\tan\beta$ or $m_\tau\tan\beta$ is large (${\cal O}(m_\t))$,
left--right
mixing can also become relevant for the sbottom and stau.
Most of the SUSY input parameters are needed to specify the
properties of the sfermions. As mentioned earlier, the effects of
mixing between the interaction and mass eigenstates are assumed
negligible for the first two generations. Furthermore, sleptons
and squarks are treated slightly differently. The physical
slepton masses $\tilde\ell_L$ and $\tilde\ell_R$ are set by
\ttt{RMSS(6)} and \ttt{RMSS(7)}. By default, the $\stau$
mixing is set by the parameters \ttt{RMSS(13)}, \ttt{RMSS(14)} and
\ttt{RMSS(17)}, which represent $M_{L_3}$, $M_{E_3}$ and $A_\tau$,
respectively, i.e.\ neither $D$-terms nor $m_\tau$ is included.
However, for \ttt{IMSS(8) = 1}, the $\stau$ masses will follow the same
pattern as for the first two generations.
Previously, it was assumed that the soft SUSY-breaking parameters
associated with the stau included $D$-terms. This is no longer the case,
and is more consistent with the treatment of the stop and sbottom.
For the first two generations of squarks,
the parameters \ttt{RMSS(8)} and \ttt{RMSS(9)} are the mass parameters
$m_{\bf 2}$ and $m_{\bf 1}$,
i.e.\ without $D$-terms included. For more generality, the
choice \ttt{IMSS(9) = 1} means that $m_{\bf 1}$ for $\tilde \u_R$
is set instead
by \ttt{RMSS(22)}, while $m_{\bf 1}$ for $\tilde \d_R$ is
\ttt{RMSS(9)}. Note that the left-handed squark mass parameters
must have the same value since they reside in the same ${\bf SU(2)_L}$
doublet. For the third generation, the parameters \ttt{RMSS(10)},
\ttt{RMSS(11)}, \ttt{RMSS(12)}, \ttt{RMSS(15)} and \ttt{RMSS(16)}
represent $M_{Q_3}$, $M_{D_3}$,
$M_{U_3}$, $A_{\b}$ and $A_{\t}$, respectively.
There is added
flexibility in the treatment of stops, sbottoms and staus.
With the flag \ttt{IMSS(5) = 1}, the properties of the third generation
sparticles can be specified by their mixing angle and mass eigenvalues
(instead of being derived from the soft SUSY-breaking parameters).
The parameters \ttt{RMSS(26) - RMSS(28)} specify the mixing angle
(in radians) for the sbottom, stop, and stau. The parameters
\ttt{RMSS(10) - RMSS(14)} specify the two stop masses, the one
sbottom mass (the other being fixed by the other parameters) and the
two stau masses. Note that the masses \ttt{RMSS(10)} and \ttt{RMSS(13)}
correspond to the left-left entries of the diagonalized matrices, while
\ttt{RMSS(11), RMSS(12)} and \ttt{RMSS(14)} correspond to the right-right
entries. These entries need not be ordered in mass.
\subsubsection{Models \label{sss:models}}
At present, the exact mechanism of SUSY breaking is unknown.
It is generally assumed that the breaking occurs spontaneously
in a set of fields that are almost entirely disconnected from
the fields of the MSSM; if SUSY is broken explicitly
in the MSSM, then some superpartners must be {\it lighter} than
the corresponding Standard Model particle, a phenomenological
disaster. The breaking of SUSY in this `hidden sector'
is then communicated to the MSSM fields through one or
several mechanisms: gravitational interactions, gauge
interactions, anomalies, etc. While any one of these
may dominate, it is also possible that all contribute at once.
We may parametrize our ignorance of the exact mechanism of SUSY
breaking by simply setting each of the soft SUSY breaking
parameters in the Lagrangian by hand. In {\Py} this approach can
be effected by setting \ttt{IMSS(1) = 1}, although some simplifications
have already been made to greatly reduce the number of parameters from
the initial more than 100.
As to specific models, several exist which predict the rich set of
measurable mass and mixing parameters from the assumed soft SUSY
breaking scenario with a much smaller set of free parameters. One
example is Supergravity (\tsc{Sugra}) inspired models, where the
number of free parameters is reduced by imposing universality at
some high scale, motivated by the apparent unification of gauge
couplings. Five parameters fixed at the gauge coupling unification
scale, $\tan\beta, M_0, m_{1/2}, A_0,$ and {\sgnmu}, are then related
to the mass parameters at the scale of Electroweak symmetry breaking
by renormalization group equations (see e.g.\ \cite{Pie97}).
The user who wants to study this and other models in detail can use
spectrum calculation programs (e.g.~\tsc{Isasusy} \cite{Bae93},
\tsc{Softsusy} \cite{All02}, \tsc{SPheno} \cite{Por03}, or
\tsc{Suspect} \cite{Djo02}), which numerically solve the renormalization
group equations (RGE) to determine the mass and mixing parameters at
the weak scale. These may then be input to {\Py} via a SLHA spectrum
file \cite{Ska03} using \ttt{IMSS(1) = 11} and
\ttt{IMSS(21)} equal to the unit number where the spectrum file has been
opened. All of {\Py}'s own internal mSUGRA machinery (see below) is then
switched off. This means that none of the other \ttt{IMSS} switches can
be used, except for \ttt{IMSS(51:53)} ($R$-parity violation),
\ttt{IMSS(10)} (force $\tilde{\chi}_2\to\tilde{\chi_1}\gamma$), and
\ttt{IMSS(11)} (gravitino is the LSP). Note that the
dependence of the $\b$ and $\t$ quark Yukawa couplings on $\tan\beta$
and the gluino mass is at present ignored when using \ttt{IMSS(1) = 11}.
As an alternative,
a run-time interface to \tsc{Isasusy} can be accessed by the options
\ttt{IMSS(1) = 12} and \ttt{IMSS(1) = 13}, in which case the \ttt{SUGRA}
routine of \tsc{Isasusy} is called by \ttt{PYINIT}. This routine then
calculates the spectrum of SUSY masses and mixings
(CP conservation, i.e.\ real-valued parameters, is assumed) and passes
the information run-time rather than in a file.
For \ttt{IMSS(1) = 12}, only the mSUGRA model of \tsc{Isasusy} can be
accessed. The mSUGRA model input parameters should then be given in
\ttt{RMSS} as for \ttt{IMSS(1) = 2}, i.e.: \ttt{RMSS(1)}$= M_{1/2}$,
\ttt{RMSS(4)} = \sgnmu, \ttt{RMSS(5)}$ = \tan\beta$,
\ttt{RMSS(8)}$ = M_0$, and \ttt{RMSS(16)}$= A_0$.
For \ttt{IMSS(1) = 13}, the full range of \tsc{Isasusy} models can be
interfaced, but the input parameters must then be given in the form of
an \tsc{Isajet} input file which \Py\ reads during initialization and
passes to \tsc{Isasusy}. The contents of the input file should be
identical to what would normally be typed when using the \tsc{Isajet}
RGE executable stand-alone (normally \ttt{isasugra.x}). The input file
should be opened by the user in his/her main program and the Logical
Unit Number should be stored in \ttt{IMSS(20)}, where {\Py} will look
for it during initialization.
The routine \ttt{PYSUGI} handles the conversion between the
conventions of {\Py} and \tsc{Isasusy}, so that conventions are
self-consistent inside \Py. In the call to \ttt{PYSUGI}, the
\ttt{RMSS} array is filled with the values produced by \tsc{Isasusy}
as for \ttt{IMSS(1) = 1}. In particular, this means that when using the
\ttt{IMSS(1) = 12} option, the mSUGRA input parameters mentioned above
will be overwritten during initialization. Cross sections and decay
widths are then calculated by \Py. Note that, since {\Py} cannot always
be expected to be linked with \tsc{Isajet}, two dummy routines and a
dummy function have been added to the {\Py} source. These are
\ttt{SUBROUTINE SUGRA}, \ttt{SUBROUTINE SSMSSM} and \ttt{FUNCTION VISAJE}.
These must first be given other names and {\Py} recompiled before proper
linking with \tsc{Isajet} can be achieved.
A problem is that the size of some \tsc{Isasusy} common blocks has
been expanded in more recent versions. Corresponding changes have been
implemented in the \ttt{PYSUGI} interface routine. Currently {\Py}
is matched to \tsc{Isajet} 7.71, and thus assumes the \ttt{SSPAR},
\ttt{SUGPAS}, \ttt{SUGMG} and \ttt{SUGXIN} common blocks to have the
forms:
\begin{verbatim}
COMMON/SSPAR/AMGLSS,AMULSS,AMURSS,AMDLSS,AMDRSS,AMSLSS
&,AMSRSS,AMCLSS,AMCRSS,AMBLSS,AMBRSS,AMB1SS,AMB2SS
&,AMTLSS,AMTRSS,AMT1SS,AMT2SS,AMELSS,AMERSS,AMMLSS,AMMRSS
&,AMLLSS,AMLRSS,AML1SS,AML2SS,AMN1SS,AMN2SS,AMN3SS
&,TWOM1,RV2V1,AMZ1SS,AMZ2SS,AMZ3SS,AMZ4SS,ZMIXSS(4,4)
&,AMW1SS,AMW2SS
&,GAMMAL,GAMMAR,AMHL,AMHH,AMHA,AMHC,ALFAH,AAT,THETAT
&,AAB,THETAB,AAL,THETAL,AMGVSS,MTQ,MBQ,MLQ,FBMA,
&VUQ,VDQ
REAL AMGLSS,AMULSS,AMURSS,AMDLSS,AMDRSS,AMSLSS
&,AMSRSS,AMCLSS,AMCRSS,AMBLSS,AMBRSS,AMB1SS,AMB2SS
&,AMTLSS,AMTRSS,AMT1SS,AMT2SS,AMELSS,AMERSS,AMMLSS,AMMRSS
&,AMLLSS,AMLRSS,AML1SS,AML2SS,AMN1SS,AMN2SS,AMN3SS
&,TWOM1,RV2V1,AMZ1SS,AMZ2SS,AMZ3SS,AMZ4SS,ZMIXSS
&,AMW1SS,AMW2SS
&,GAMMAL,GAMMAR,AMHL,AMHH,AMHA,AMHC,ALFAH,AAT,THETAT
&,AAB,THETAB,AAL,THETAL,AMGVSS,MTQ,MBQ,MLQ,FBMA,VUQ,VDQ
COMMON /SUGPAS/ XTANB,MSUSY,AMT,MGUT,MU,G2,GP,V,VP,XW,
&A1MZ,A2MZ,ASMZ,FTAMZ,FBMZ,B,SIN2B,FTMT,G3MT,VEV,HIGFRZ,
&FNMZ,AMNRMJ,NOGOOD,IAL3UN,ITACHY,MHPNEG,ASM3,
&VUMT,VDMT,ASMTP,ASMSS,M3Q
REAL XTANB,MSUSY,AMT,MGUT,MU,G2,GP,V,VP,XW,
&A1MZ,A2MZ,ASMZ,FTAMZ,FBMZ,B,SIN2B,FTMT,G3MT,VEV,HIGFRZ,
&FNMZ,AMNRMJ,ASM3,VUMT,VDMT,ASMTP,ASMSS,M3Q
INTEGER NOGOOD,IAL3UN,ITACHY,MHPNEG
COMMON /SUGMG/ MSS(32),GSS(31),MGUTSS,GGUTSS,AGUTSS,FTGUT,
&FBGUT,FTAGUT,FNGUT
REAL MSS,GSS,MGUTSS,GGUTSS,AGUTSS,FTGUT,FBGUT,FTAGUT,FNGUT
COMMON /SUGXIN/ XISAIN(24),XSUGIN(7),XGMIN(14),XNRIN(4),
&XAMIN(7)
REAL XISAIN,XSUGIN,XGMIN,XNRIN,XAMIN
\end{verbatim}
\tsc{Isasusy} users are warned to check that no incompatibilities arise
between the versions actually used. Unfortunately there is no universal
solution to this problem: the Fortran standard does not allow you
dynamically to vary the size of a (named) common block. So if you use an
earlier \tsc{Isasusy} version, you have to shrink the size accordingly,
and for a later you may have to check that the above common blocks
have not been expanded further.
As a cross check, the option \ttt{IMSS(1) = 2}
uses approximate analytical solutions of the renormalization
group equations \cite{Dre95}, which reproduce the output of
\tsc{Isasusy} within $\simeq 10\%$ (based on comparisons
of masses, decay widths, production cross sections, etc.).
This option is intended for debugging only, and does not represent
the state-of-the-art.
In \tsc{Sugra} and in other models with the SUSY breaking scale of order
$M_{\mrm{GUT}}$, the spin--3/2 superpartner
of the graviton, the gravitino $\gravitino$ (code 1000039), has
a mass of order $M_\W$ and interacts only gravitationally.
In models of gauge-mediated SUSY breaking \cite{Din96}, however,
the gravitino can play a crucial role in the phenomenology,
and can be the lightest superpartner (LSP). Typically, sfermions decay
to fermions and gravitinos, and neutralinos, chargino, and gauginos
decay to gauge or Higgs bosons and gravitinos.
Depending on the gravitino mass, the decay lengths can be substantial
on the scale of colliders. {\Py} correctly handles finite decay lengths
for all sparticles.
$R$-parity is a possible symmetry of the SUSY Lagrangian that prevents
problems of rapid proton decay and allows for a viable dark matter
candidate. However, it is also possible to allow a restricted amount of
$R$-parity violation. At present, there is no theoretical consensus that
$R$-parity should be conserved, even in string models.
In the production of superpartners, {\Py} assumes
$R$-parity conservation (at least on the time and distance
scale of a typical collider experiment), and only lowest order,
sparticle pair production processes are included. Only those
processes with $\e^+\e^-$, $\mu^+\mu^-$, or quark and gluon initial
states are simulated. Tables~\ref{t:procfive}, \ref{t:procsix} and
\ref{t:procseven} list available SUSY processes. In processes
210 and 213, $\sell$ refers to both $\se$ and $\smu$. For ease of
readability, we have removed the subscript $L$ on $\snu$.
$\tp_i\tm_i, \stau_i\stau_j^*$ and $\stau_i\snu_{\tau}^*$
production correctly account for sfermion mixing. Several processes
are conspicuously absent from the table. For example, processes
\ttt{255} and \ttt{257} would simulate the associated production
of right-handed squarks with charginos. Since the right-handed squark
only couples to the higgsino component of the chargino, the interaction
strength is proportional to the quark mass, so these processes can
be ignored.
By default, only $R$-parity conserving decays are allowed, so that one
sparticle is stable, either the lightest neutralino, the gravitino,
or a sneutrino. SUSY decays of the top quark are included, but all
other SM particle decays are unaltered.
Generally, the decays of the superpartners are calculated using the
formulae of refs.~\cite{Gun88,Bar86a,Bar86b,Bar95}.
All decays are spin averaged. Decays involving $\sbo$ and $\st$
use the formulae of \cite{Bar95}, so they are valid for large values of
$\tan\beta$. The one loop decays $\chio_j\to\chio_i\gamma$ and
$\st\to \c\chio_1$ are also included, but only with approximate formula.
Typically, these decays are only important when other decays are not
allowed because of mixing effects or phase space considerations.
One difference between the SUSY simulation and the other parts of
the program is that it is not beforehand known which sparticles
may be stable. Normally this would mean either the $\chio^0_1$
or the gravitino $\grav$, but in principle also other sparticles could
be stable. The ones found to be stable have their \ttt{MWID(KC)} and
\ttt{MDCY(KC,1)} values set zero at initialization. If several
\ttt{PYINIT} calls are made in the same run, with different SUSY
parameters, the ones set zero above are not necessarily set
back to nonzero values, since most original values are not saved
anywhere. As an exception to this rule, the \ttt{PYMSIN} SUSY
initialization routine, called by \ttt{PYINIT}, does save and restore
the \ttt{MWID(KC)} and \ttt{MDCY(KC,1)} values of the lightest
SUSY particle. It is therefore possible to combine several \ttt{PYINIT}
calls in a single run, provided that only the lightest SUSY particle is
stable. If this is not the case, \ttt{MWID(KC)} and \ttt{MDCY(KC,1)}
values may have to be reset by hand, or else some particles that ought
to decay will not do that.
\subsubsection{SUSY examples}
The SUSY routines and common-block variables are described in
section \ref{ss:susycode}. To illustrate the usage of the switches and
parameters, we give six simple examples.
\textit{Example 1: Light Stop}\\
The first example is an MSSM model with a light neutralino $\chio_1$
and a light stop $\st_1$, so that $\t \to \st_1\chio_1$ can occur.
The input parameters are\\
\ttt{IMSS(1) = 1}, \ttt{RMSS(1) = 70.}, \ttt{RMSS(2) = 70.},
\ttt{RMSS(3) = 225.}, \ttt{RMSS(4) = -40.},\\
\ttt{RMSS(5) = 1.5}, \ttt{RMSS(6) = 100.}, \ttt{RMSS(7) = 125.},
\ttt{RMSS(8) = 250.},\\
\ttt{RMSS(9) = 250.}, \ttt{RMSS(10) = 1500.}, \ttt{RMSS(11) = 1500.},
\ttt{RMSS(12) = -128.},\\
\ttt{RMSS(13) = 100.}, \ttt{RMSS(14) = 125.}, \ttt{RMSS(15) = 800.},
\ttt{RMSS(16) = 800.},\\
\ttt{RMSS(17) = 0.}, and \ttt{RMSS(19) = 400.0.}\\
The top mass is fixed at 175 GeV, \ttt{PMAS(6,1) = 175.0}.
The resulting model has $M_{\st_1} = 55$ GeV and $M_{\chio_1} = 38$ GeV.
\ttt{IMSS(1) = 1} turns on the MSSM simulation.
By default, there are no intrinsic relations between the gaugino masses,
so $M_1 = 70$ GeV, $M_2 = 70$ GeV, and $M_3 = 225$ GeV. The pole mass of
the gluino is slightly higher than the parameter $M_3$, and the
decay $\glu \to\st_1^*\t+\st_1\tbar$ occurs almost 100\% of the time.
\textit{Example 2: SUSY Les Houches Accord spectrum}\\
The second example shows how to input a spectrum file in the SUSY Les
Houches Accord format \cite{Ska03}
to {\Py}. First, you should set \ttt{IMSS(1) = 11} and
open the spectrum file you want to use on some unused Logical Unit Number.
Then, set \ttt{IMSS(21)} equal to that number, to tell {\Py} where to read
the spectrum file from. This should be done somewhere in your main program
before calling \ttt{PYINIT}. During the call to \ttt{PYINIT}, {\Py} will read
the spectrum file, perform a number of consistency checks and issue
warning messages if it finds something it does not understand or which seems
inconsistent. E.g.~\ttt{BLOCK GAUGE} will normally be present in the
spectrum file, but since {\Py} currently cannot
use the information in that block, it will issue a warning that the block
will be ignored. In case a decay table is also desired to be read in, the
Logical Unit Number on which the decay table is opened should be put in
\ttt{IMSS(22)}. To avoid inconsistencies, the spectrum and the
decay table should normally go together, so \ttt{IMSS(22)} should normally be
equal to \ttt{IMSS(21)}.
\textit{Example 3: Calling \tsc{Isasusy 7.71} at runtime using
\ttt{IMSS(1) = 12}}\\
The third example shows how to use the built-in run-time interface to
\tsc{Isasusy} with the \ttt{IMSS(1) = 12} option.
First, the {\Py} source code needs to be changed. Rename the function
\ttt{VISAJE} to, for example, \ttt{FDUMMY}, rename the subroutines
\ttt{SUGRA} and \ttt{SSMSSM} to e.g.\ \ttt{SDUMM1} and \ttt{SDUMM2},
and recompile. In the calling program, set \ttt{IMSS(1) = 12} and the
\ttt{RMSS} input parameters exactly as in example 5, and compile the
executable while linked to both \tsc{Isajet} and the modified \Py.
The resulting mass and mixing spectrum is printed in the {\Py} output.
\textit{Example 4: Calling \tsc{Isasusy 7.71} at runtime using
\ttt{IMSS(1) = 13}}\\
The fourth example shows how to use the built-in run-time interface to
\tsc{Isasusy} with the \ttt{IMSS(1) = 13} option. First, the {\Py}
source code needs to be changed, cf.\ the previous example. In the
calling program, set \ttt{IMSS(1) = 13} and open an \tsc{Isajet} SUSY
model input file on any available Logical Unit Number. The contents of
the file should be exactly identical to what would normally be typed
when using the \tsc{Isajet} RGE executable stand-alone (normally
\ttt{isasugra.x}). Then, store that Unit Number in \ttt{IMSS(20)},
that will enable {\Py} to access the correct file during
initialization. Compile the executable while linked to both
\tsc{Isajet} and the modified \Py. The resulting mass and mixing
spectrum is printed in the {\Py} output.
\textit{Example 5: Approximate \tsc{SUGRA}}\\
This example shows you how to get a (very) approximate SUGRA model.
Note that this way of obtaining the SUSY spectrum should
never be used for serious studies. The input parameters are\\
\ttt{IMSS(1) = 2}, \ttt{RMSS(1) = 200.}, \ttt{RMSS(4) = 1.},
\ttt{RMSS(5) = 10.}, \ttt{RMSS(8) = 800.}, and
\ttt{RMSS(16) = 0.0}.\\
The resulting model has
$M_{\sd_L}=901$ GeV, $M_{\su_R}=890$ GeV, $M_{\st_1}=538$ GeV,
$M_{\se_L}=814$ GeV, $M_{\glu}=560$ GeV, $M_{\chio_1}=80$ GeV,
$M_{\chip_1}=151$ GeV, $M_{\hrm}=110$ GeV, and $M_{A}=883$ GeV. It
corresponds to the choice $M_0$=800 GeV, $M_{1/2}=$200 GeV,
$\tan\beta=10$, $A_0=0$, and \sgnmu $> 0$. The output is similar
to an \tsc{Isasusy} run, but there is not exact agreement.
\textit{Example 6: \tsc{Isasusy 7.71} Model}\\
The final example demonstrates how to convert the output of
an \tsc{Isasusy} run directly into the {\Py} format, i.e.\ if
SLHA output is not available.
This assumes that you already made an \tsc{Isasusy} run, e.g. with the
equivalents of the input parameters above. From the output of this run
you can now extract those physical parameters that need to be handed to
{\Py}, in the above example\\
\ttt{IMSS(1) = 1}, \ttt{IMSS(3) = 1}, \ttt{IMSS(8) = 0},
\ttt{IMSS(9) = 1}, \ttt{RMSS(1) = 79.61},\\
\ttt{RMSS(2) = 155.51}, \ttt{RMSS(3) = 533.1},
\ttt{RMSS(4) = 241.30}, \ttt{RMSS(5) = 10.}, \\
\ttt{RMSS(6) = 808.0}, \ttt{RMSS(7) = 802.8},
\ttt{RMSS(8) = 878.4}, \ttt{RMSS(9) = 877.1},\\
\ttt{RMSS(10) = 743.81}, \ttt{RMSS(11) = 871.26},
\ttt{RMSS(12) = 569.87}, \ttt{RMSS(13) = 803.20},\\
\ttt{RMSS(14) = 794.71}, \ttt{RMSS(15) = -554.96},
\ttt{RMSS(16) = -383.23}, \ttt{RMSS(17) = -126.11},\\
\ttt{RMSS(19) = 829.94} and \ttt{RMSS(22) = 878.5}.
\subsubsection{$R$-Parity Violation}
\label{sss:rparityviol}
$R$-parity, defined as $R=(-1)^{2S+3B+L}$, is a discrete multiplicative
symmetry where $S$ is the particle spin, $B$ is the baryon number,
and $L$ is the lepton number. All SM particles have $R=1$, while
all superpartners have $R=-1$, so a single SUSY particle cannot
decay into just SM particles if
$R$-parity is conserved. In this case, the lightest superpartner (LSP) is
absolutely stable. Astrophysical considerations imply that a stable LSP
should be electrically neutral. Viable candidates are the lightest
neutralino, the lightest sneutrino, or alternatively the gravitino. Since
the LSP can carry away energy without interacting in a detector, the apparent
violation of momentum conservation is an important part of SUSY
phenomenology. Also, when $R$-parity is conserved, superpartners must be
produced in pairs from a SM initial state. The breaking of the $R$-parity
symmetry would result in lepton- and/or baryon-number-violating processes.
While there are strong experimental constraints on some classes of
$R$-parity-violating interactions, others are hardly constrained at all.
One simple extension of the MSSM is to break the multiplicative
$R$-parity symmetry.
Presently, neither experiment nor any theoretical argument
demand $R$-parity conservation, so it is natural to consider the
most general case of $R$-parity breaking.
It is convenient to introduce a function of superfields
called the superpotential, from which the Feynman rules for
$R$-parity-violating processes can be derived.
The $R$-parity-violating (RPV) terms
which can contribute to the superpotential are:
\begin{equation}
W_{RPV} = \lambda_{ijk} L^i L^j \bar{E}^k +
\lambda^{'}_{ijk} L^i Q^j \bar{D}^k +
\lambda^{''}_{ijk} \bar{U}^i \bar{D}^j \bar{D}^k +
\epsilon_i L_iH_2
\label{eq:superpot}
\end{equation}
where $i,j,k$ are generation indices (1,2,3), $L^i_1 \equiv \nu^i_L$,
$L^i_2=\ell^i_L$ and $Q^i_1=u^i_{L}$, $Q^i_2=d^i_{L}$ are lepton and quark
components of ${\bf SU(2)_L}$ doublet superfields, and $E^i=e^i_{R}$,
$D^i=d^i_{R}$ and $U^i=u^i_R$ are lepton, down- and up-quark
${\bf SU(2)_L}$ singlet superfields, respectively. The unwritten
${\bf SU(2)_L}$ and ${\bf SU(3)_C}$
indices imply that the first term is antisymmetric under $i \leftrightarrow
j$, and the third term is antisymmetric under $j \leftrightarrow
k$. Therefore, $i \neq j$ in $L^i L^j \bar{E}^k$ and $j \neq k$ in $\bar{U}^i
\bar{D}^j \bar{D}^k$. The coefficients $\lambda_{ijk}$, $\lambda^{'}_{ijk}$,
$\lambda^{''}_{ijk}$, and $\epsilon_i$ are Yukawa couplings, and there is no
{\it a priori} generic prediction for their values. In principle, $W_{RPV}$
contains 48 extra parameters over the $R$-parity-conserving MSSM case. In
{\Py} the effects of the last term in eq.~(\ref{eq:superpot}) are not
included.
Expanding eq.~(\ref{eq:superpot})
as a function of the superfield components, the interaction
Lagrangian derived from the first term is
\begin{equation}
{\cal{L}}_{LLE} = \lambda_{ijk} \left\{ \tilde{\nu}_L^i e_L^j \bar{e}^k_R +
\tilde{e}_L^i \nu_L^j \bar{e}^k_R +
(\tilde{e}_R^k)^* \nu_L^i e^j_L + h.c. \right\}
\label{eq:RPVLLE}
\end{equation}
and from the second term,
\begin{eqnarray}
{\cal{L}}_{LQD} = \lambda^{'}_{ijk} \left\{
\tilde{\nu}_L^i d_L^j \bar{d}^k_R -
\tilde{e}_L^i u_L^j \bar{d}^k_R +
\tilde{d}_L^j \nu_L^i \bar{d}^k_R -
\tilde{u}_L^j e_L^i \bar{d}^k_R + \right. \nonumber \\
\left. (\tilde{d}_R^k)^* \nu_L^i d^j_L -
(\tilde{d}_R^k)^* e_L^i u^j_L + h.c. \right\}
\end{eqnarray}
Both of these sets of interactions violate lepton number.
The $\bar U\bar D\bar D$ term, instead, violates
baryon number.
In principle, all types of $R$-parity-violating terms may co-exist,
but this can lead to a proton
with a lifetime shorter than the present experimental limits.
The simplest way to avoid this
is to allow only operators which conserve baryon-number but violate
lepton-number or vice versa.
There are several effects on the SUSY phenomenology due to these
new couplings: (1) lepton- or baryon-number-violating processes
are allowed, including the
production of single sparticles (instead of pair production),
(2) the LSP is no longer stable, but
can decay to SM particles within a collider detector,
and
(3) because it is unstable,
the LSP need not be the neutralino or sneutrino, but
can be charged and/or coloured.
In the current version of {\Py}, decays of supersymmetric particles
to SM particles via two different types of lepton-number-violating
couplings and one type of baryon-number-violating couplings
can be invoked \cite{Ska01,Sjo03}.
Complete matrix elements (including $L-R$ mixing for all sfermion
generations) for all two-body sfermion and three-body neutralino,
chargino, and gluino decays are included
(as given in \cite{Dre00}). The final-state fermions are treated as
massive in the phase space integrations and in the matrix elements
for $\b$, $\t$, and $\tau$.
The existence of $R$-odd couplings also allows for single sparticle
production, i.e.\ there is no requirement that SUSY particles should be
produced in pairs. Single sparticle production cross sections are not yet
included in the program, and it may require some rethinking of the parton
shower to do so. For low-mass sparticles, the
associated error is estimated to be negligible, as long as the $R$-violating
couplings are smaller than the gauge couplings. For higher-mass sparticles,
the reduction of the
phase space for pair production becomes an important factor, and single
sparticle production could dominate even for very small values of the
$R$-violating couplings. The total SUSY
production cross sections, as calculated by {\Py} in its current form are
thus underestimated, possibly quite severely for heavy-mass sparticles.
Three possibilities exist for the initializations of the couplings,
representing a fair but not exhaustive range of models. The
first, selected by setting \ttt{IMSS(51) = 1} for LLE, \ttt{IMSS(52) = 1} for
LQD, and/or \ttt{IMSS(53) = 1} for UDD type couplings, sets all the couplings,
independent of generation, to a
common value of $10^{-\mbox{\scriptsize\ttt{RMSS(51)}}}$,
$10^{-\mbox{\scriptsize\ttt{RMSS(52)}}}$, and/or
$10^{-\mbox{\scriptsize\ttt{RMSS(53)}}}$, depending on which couplings are
activated.
Taking now LLE couplings as an example, setting
\ttt{IMSS(51) = 2} causes the LLE couplings
to be initialized (in \ttt{PYINIT}) to so-called `natural'
generation-hierarchical values, as proposed in \cite{Hin93}. These values,
inspired by the structure of the Yukawa couplings in the SM, are defined by:
\begin{equation}
\begin{array}{rcl}
|\lambda_{ijk}|^2 & = & (\mbox{\ttt{RMSS(51)}})^
2\hat{m}_{e_i}\hat{m}_{e_j}\hat{m}_{e_k}\\
|\lambda'_{ijk}|^2 & = &
(\mbox{\ttt{RMSS(52)}})^ 2 \hat{m}_{e_i} \hat{m}_{q_j}
\hat{m}_{d_k}\\
|\lambda''_{ijk}|^2 & = & (\mbox{\ttt{RMSS(53)}})^
2\hat{m}_{q_i}\hat{m}_{q_j}\hat{m}_{q_k}\end{array} \hspace*{1cm} ;
\hspace*{0.7cm} \hat{m}\equiv
\frac{m}{v} = \frac{m}{126\mrm{GeV}}
\label{eq:rvnatval}
\end{equation}
where $m_{q_i}$ is the arithmetic mean of $m_{u_i}$ and $m_{d_i}$.
The third option available is to set \ttt{IMSS(51) = 3}, \ttt{IMSS(52) = 3},
and/or \ttt{IMSS(53) = 3},
in which case all the relevant couplings are zero by default (but the
corresponding lepton- or baryon-number-violating processes are turned on)
and the user is expected to enter the non-zero coupling values by hand.
(Where antisymmetry is required, half of the entries are automatically
derived from the other half, see \ttt{IMSS(51) = 3} and \ttt{IMSS(53) = 3}.)
\ttt{RVLAM($i$,$j$,$k$)} contains the $\lambda_{ijk}$,
\ttt{RVLAMP($i$,$j$,$k$)} contains the $\lambda'_{ijk}$ couplings, and
\ttt{RVLAMB($i$,$j$,$k$)} contains the $\lambda''_{ijk}$ couplings.
\subsubsection{NMSSM}
In the Next-to-Minimal Supersymmetric Standard Model (NMSSM), three new
particles appear: a new CP-even Higgs boson $\H^0_3$ (code 45),
a new CP-odd Higgs boson $\A^0_2$ (code 46), and an additional
neutralino $\chi_5^0$ (code 1000045), where the particle codes are
the ones tentatively adopted by the SUSY Les Houches Accord (SLHA)
community \cite{All06}.
{\Py} does not contain any internal machinery for doing calculations
in the NMSSM. Thus, the basic scattering processes should be generated
by an external program (e.g.\ \tsc{CompHEP}/\tsc{CalcHEP} \cite{Puk99})
and handed to {\Py} via the Les Houches Accord interface for parton-level
events (LHA). This should then be combined with either setting the NMSSM
resonance decays by hand, or by reading in an SLHA decay table prepared
by an external decay package (e.g.\ \tsc{NMHDecay} \cite{Ell05}). One
possible chain of steps for generating fully simulated events for the
NMSSM,
starting from a high-scale model definition,
would thus be (see e.g.\ \cite{Puk05}):
\begin{Enumerate}
\item Obtain the EW scale masses and couplings for the model (e.g.\ by
running \tsc{NMHDecay}), and store the results as a SLHA spectrum file.
\item Compute decay widths and branching ratios for all relevant
particles (by hand or using some code), and store the resulting numbers
in the SLHA decay table format.
\item Pass the spectrum to an NMSSM Matrix Element level generator
(e.g.\ \tsc{CompHEP}/\tsc{CalcHEP}), and obtain a set of elementary
$2 \to 2$ (or $2 \to$ `a few') scatterings.
\item Read in the SLHA spectrum and decay table into {\Py} using
\ttt{IMSS(1) = 11} and \ttt{IMSS(13) = 1} (you need to set \ttt{IMSS(21)}
and possibly \ttt{IMSS(22)} as well).
\item Read in the ME level events into {\Py} using the LHA interface
routines \ttt{UPINIT} and \ttt{UPEVNT}.
\end{Enumerate}
\subsubsection{Long-lived coloured sparticles}
In SUSY scenarios, the coloured sparticles --- squarks and gluinos
--- typically have widths of several GeV, and thus decay well before
they would have had time to hadronize. There are specific cases in
which one of them is more long-lived or even (quasi-)stable, however.
A recent example is the split SUSY scenarios, wherein the gauginos
are rather light and the sfermions very heavy \cite{Ark05}.
Then the gluino decay is strongly suppressed, since it has to go
via a virtual squark. In such cases, hadronic states may have time
to form around them. These are called $R$-hadrons, since they carry
negative $R$-parity.
In the long-lived-gluino case, the set of possible $R$-hadrons
include `gluino-balls' $\sg\g$, `gluino-mesons' $\sg\q\qbar$
and `gluino-baryons' $\sg\q\q\q$, in multiplets similarly to normal
hadrons, but of course with differences in the mass spectra and other
properties. In the case of a long-lived squark, such as the stop,
there would be `squark-mesons' $\sq\qbar$ and `squark-baryons'
$\sq\q\q$. When they pass through a detector, they may undergo
charge- and baryon-number-changing interactions \cite{Kra04},
giving rise to quite spectacular and characteristic experimental
signals \cite{Kra04a}.
Owing to the somewhat special nature of their production, these
$R$-hadrons do not fit exactly into the current {\Py} event generation
chain, although all the separate tools exist. Therefore the
production of gluino- and stop-hadrons is simulated by two separate
add-on programs, available on the {\Py} webpage, where the calling
sequence is modified as appropriate. In future versions with a
modified administrative structure, the intention is to include this
capability in the standard distribution.
Finally, we note that a numbering scheme for $R$-hadrons is under
development, for eventual inclusion in the PDG standard. The
abovementioned programs comply with the draft proposal, but the
absence of an approved standard is another reason why the programs
have not yet been fully integrated into {\Py}.
\subsection{Polarization}
\label{ss:polarization}
In most processes, incoming beams are assumed unpolarized. However,
especially for $\e^+\e^-$ linear collider studies, polarized beams
would provide further important information on many new physics
phenomena, and at times help to suppress backgrounds. Therefore a
few process cross sections are now available
also for polarized incoming beams. The average polarization of the
two beams is then set by \ttt{PARJ(131)} and \ttt{PARJ(132)},
respectively. In some cases, noted below, \ttt{MSTP(50)} need also
be switched on to access the formulae for polarized beams.
Process 25, $\W^+ \W^-$ pair production, allows polarized incoming lepton
beam particles. The polarization effects are included both in the production
matrix elements and in the angular distribution of the final four fermions.
Note that the matrix element used \cite{Mah98} is for on-shell $\W$
production, with a suppression factor added for finite width effects.
This polarized cross section expression, evaluated at vanishing polarization,
disagrees with the standard unpolarized one, which presumably is the more
accurate of the two. The difference can be quite significant below threshold
and at very high energies. This can be traced to the simplified description
of off-shell $\W$'s in the polarized formulae. Good agreement is obtained
either by switching off the $\W$ width with \ttt{MSTP(42) = 0} or by
restricting the $\W$ mass ranges (with \ttt{CKIN(41) - CKIN(44)}) to be
close to on-shell. It is therefore necessary to set \ttt{MSTP(50) = 1}
to switch from the default standard unpolarized formulae to the polarized
ones.
Also many SUSY production processes now include the effects from
polarization of the incoming fermion beams. This applies for scalar pair
production, with the exception of sneutrino pair production and
$\hrm^0 \A^0$ and $\H^0 \A^0$ production, this omission being an oversight
at the time of this release, but easily remedied in the future.
The effect of polarized photons is included in the process
$\gamma \gamma \to \F_k \Fbar_k$, process 85. Here the array values PARJ(131)
and PARJ(132) are used to define the average longitudinal polarization of
the two photons.
\subsection{Main Processes by Machine}
In the previous section we have already commented on which processes
have limited validity, or have different meanings (according to
conventional terminology) in different contexts. Let us just repeat
a few of the main points to be remembered for different machines.
\subsubsection{$\ee$ collisions}
The main annihilation process is number 1, $\ee \to \Z^0$,
where in fact the full $\gamma^*/\Z^0$ interference structure is
included. This process can be used, with some confidence, for
c.m.\ energies from about 4 GeV upwards, i.e.\ at
DORIS/CESR/PEP-II/KEKB, PETRA/PEP, TRISTAN, LEP,
and any future linear colliders.
(To get below 10 GeV, you have to change \ttt{PARP(2)}, however.)
This is the default process obtained when \ttt{MSEL = 1}, i.e.\
when you do not change anything yourself.
Process 141 contains a $\Z'^0$, including
full interference with the standard $\gammaZ$. With the value
\ttt{MSTP(44) = 4} in fact one is back at the standard $\gammaZ$
structure, i.e.\ the $\Z'^0$ piece has been switched off. Even so,
this process may be useful, since it can simulate e.g.
$\ee \to \hrm^0 \A^0$. Since the $\hrm^0$ may in its turn decay to
$\Z^0 \Z^0$, a decay channel of the ordinary $\Z^0$ to
$\hrm^0 \A^0$, although physically correct, would be technically
confusing. In particular, it would be messy to set the original
$\Z^0$ to decay one way and the subsequent ones another. So, in
this sense, the $\Z'^0$ could be used as a copy of the ordinary
$\Z^0$, but with a distinguishable label.
The process $\ee \to \Upsilon$ does not exist as a separate process
in {\Py}, but can be simulated by using \ttt{PYONIA}, see section
\ref{ss:oniadecays}.
At LEP 2 and even higher energy machines, the simple $s$-channel
process 1 loses out to other processes, such as
$\ee \to \Z^0 \Z^0$ and $\ee \to \W^+ \W^-$, i.e.\ processes
22 and 25. The former process in fact includes the structure
$\ee \to (\gammaZ)(\gammaZ)$, which means that the cross section
is singular if either of the two $\gammaZ$ masses is allowed to
vanish. A mass cut therefore needs to be introduced, and is
actually also used in other processes, such as $\ee \to \W^+ \W^-$.
For practical
applications, both with respect to cross sections and to event
shapes, it is imperative to include initial-state radiation effects.
Therefore \ttt{MSTP(11) = 1} is the default, wherein exponentiated
electron-inside-electron distributions are used to give the
momentum of the actually interacting electron. By radiative
corrections to process 1, such processes as $\ee \to \gamma \Z^0$
are therefore automatically generated. If process 19 were to be
used at the same time, this would mean that radiation were to be
double-counted. In the alternative \ttt{MSTP(11) = 0}, electrons are
assumed to deposit their full energy in
the hard process, i.e.\ initial-state QED radiation is not included.
This option is very useful, since it often corresponds to the
`ideal' events that one wants to correct back to.
Resolved electrons also means that one may have interactions
between photons. This opens up the whole field of $\gamma\gamma$
processes, which is described in section \ref{ss:photoanddisclass}.
In particular, with \ttt{'gamma/e+','gamma/e-'} as beam and target
particles in a \ttt{PYINIT} call, a flux of photons of different
virtualities is convoluted with a description of direct and resolved
photon interaction processes, including both low-$\pT$ and
high-$\pT$ processes. This machinery is directed to the description
of the QCD processes, and does e.g.\ not address the production of
gauge bosons or other such particles by the interactions of resolved
photons. For the latter kind of applications, a simpler description
of partons inside photons inside electrons may be obtained with the
\ttt{MSTP(12) = 1} options and $\e^{\pm}$ as beam and target particles.
The thrust of the {\Py} programs is towards processes that involve
hadron production, one way or another. Because of generalizations
from other areas, also a few completely
non-hadronic processes are available. These include Bhabha
scattering, $\ee \to \ee$ in process 10, and photon pair production,
$\ee \to \gamma \gamma$ in process 18. However, note that the
precision that could be expected in a {\Py} simulation of those
processes is certainly far less than that of dedicated programs.
For one thing, electroweak loop effects are not included.
For another, nowhere is the electron mass taken into account,
which means that explicit cut-offs at some minimum $\pT$ are always
necessary.
\subsubsection{Lepton--hadron collisions}
The main option for photoproduction and Deeply Inelastic Scattering
(DIS) physics is provided by the {\galep} option as beam
or target in a \ttt{PYINIT} call, see section
\ref{ss:photoanddisclass}. The $Q^2$ range to be covered, and other
kinematics constraints, can be set by \ttt{CKIN} values. By default,
when the whole $Q^2$ range is populated, obviously photoproduction
dominates.
The older DIS process 10, $\ell \q \to \ell' \q'$, includes
$\gamma^0/\Z^0/\W^{\pm}$ exchange, with full interference, as
described in section \ref{sss:DISclass}. The $\Z^0/\W^{\pm}$
contributions are not implemented in the {\galep}
machinery. Therefore process 10 is still the main option for
physics at very high $Q^2$, but has been superseded for lower $Q^2$.
Radiation off the incoming lepton leg is included by \ttt{MSTP(11) = 1}
and off the outgoing one by \ttt{MSTJ(41) = 2} (both are default). Note
that both QED and QCD radiation (off the $\e$ and the $\q$ legs,
respectively) are allowed to modify the $x$ and $Q^2$ values of the
process, while the conventional approach in the literature is to allow
only the former. Therefore an option (on by default) has been added to
preserve these values by a post-facto rescaling, \ttt{MSTP(23) = 1}.
Further comments on HERA applications are found in
\cite{Sjo92b,Fri00}.
\subsubsection{Hadron--hadron collisions}
The default is to include QCD jet production by $2 \to 2$ processes,
see section \ref{sss:QCDjetclass}. Since the differential
cross section is divergent for $\pT \to 0$, a lower cut-off has to
be introduced. Normally that cut-off is given by the user-set
$\pTmin$ value in \ttt{CKIN(3)}. If \ttt{CKIN(3)} is chosen
smaller than a given value of the order of 2 GeV (see \ttt{PARP(81)}
and \ttt{PARP(82)}), then low-$\pT$ events are also switched on.
The jet cross section is regularized at low $\pT$, so as to
obtain a smooth joining between the high-$\pT$ and the low-$\pT$
descriptions, see further section \ref{ss:multint}.
As \ttt{CKIN(3)} is varied, the jump from one
scenario to another is abrupt, in terms of cross section: in a
high-energy hadron collider, the cross section for jets down to
a $\pTmin$ scale of a few GeV can well reach values much
larger than the total inelastic, non-diffractive cross section.
Clearly this is nonsense; therefore either $\pTmin$ should
be picked so large that the jet cross section be only a fraction
of the total one, or else one should select $\pTmin = 0$ and
make use of the full description.
If one switches to \ttt{MSEL = 2}, also elastic and diffractive
processes are switched on, see section \ref{sss:minbiasclass}.
However, the simulation of these processes is fairly primitive,
and should not be used for dedicated studies,
but only to estimate how much they may contaminate the class of
non-diffractive minimum-bias events.
Most processes can be simulated in hadron colliders, since the
bulk of {\Py} processes can be initiated by quarks or gluons.
However, there are limits. Currently we include no photon
or lepton parton distributions, which means that a process like
$\gamma \q \to \gamma \q$ is not accessible. Further, the
possibility of having $\Z^0$ and $\W^{\pm}$ interacting in
processes such as 71--77 has been hardwired process by process,
and does not mean that there is a generic treatment of $\Z^0$ and
$\W^{\pm}$ distributions.
The emphasis in the hadron--hadron process description is on high
energy hadron colliders. The program can be used also at
fixed-target energies, but the multiple interaction model for
underlying events then may break down and has to be used with
caution. The limit of `safe' applicability is somewhere at around
100 GeV. Only with the simpler model obtained for \ttt{MSTP(82) = 1}
can one go arbitrarily low.
\clearpage
\section{The Process Generation Program Elements}
In the previous two sections, the physics processes and the
event-generation schemes of {\Py} have been presented. Here, finally,
the event-generation routines and the common-block variables are
described. However, routines and variables related to initial- and
final-state showers, beam remnants and underlying events, and
fragmentation and decay are relegated to subsequent sections on
these topics.
In the presentation in this section, information less important for
an efficient use of {\Py} has been put closer to the end.
We therefore begin with the main event generation routines, and
follow this by the main common-block variables.
It is useful to distinguish three phases in a normal run with {\Py}.
In the first phase, the initialization, the general character of the
run is determined. At a minimum, this requires the specification of the
incoming hadrons and the energies involved. At the discretion of the
user, it is also possible to select specific final states, and to make
a number of decisions about details in the subsequent generation.
This step is finished by a \ttt{PYINIT} call, at which time several
variables are initialized in accordance with the values set. The
second phase consists of the main loop over the number of events,
with each new event being generated by a call to either \ttt{PYEVNT} or
\ttt{PYEVNW} (depending on which underlying-event and parton-shower
framework is desired; below we shall often not make the distinction
explicit, referring to both routines by \ttt{PYEVNT} generically).
This event
may then be analysed, using information stored in some common blocks,
and the statistics accumulated. In the final phase, results are
presented. This may often be done without the invocation of any {\Py}
routines. From \ttt{PYSTAT}, however, it is possible to obtain a
useful list of cross sections for the different subprocesses.
\subsection{The Main Subroutines}
\label{ss:PYTmainroutines}
There are two routines that you must know: \ttt{PYINIT} for
initialization, and either \ttt{PYEVNT} or \ttt{PYEVNW} for the
subsequent generation of each new event.
In addition, the cross section and other kinds of
information available with \ttt{PYSTAT} are frequently useful. The
other two routines described here, \ttt{PYFRAM} and \ttt{PYKCUT},
are of more specialized interest.
\drawbox{CALL PYINIT(FRAME,BEAM,TARGET,WIN)}\label{p:PYINIT}
\begin{entry}
\itemc{Purpose:} to initialize the generation procedure. Normally it
is foreseen that this call will be followed by many \ttt{PYEVNT}
(or \ttt{PYEVNW}) ones,
to generate a sample of the event kind specified by the \ttt{PYINIT}
call. (For problems with cross section estimates in runs of very few
events per \ttt{PYINIT} call, see the description for \ttt{PYSTAT(1)}
below in this section.)
\iteme{FRAME :} a character variable used to specify the frame of the
experiment. Upper-case and lower-case letters may be freely mixed.
\begin{subentry}
\iteme{= 'CMS' :} colliding beam experiment in c.m.\ frame, with beam
momentum in $+z$ direction and target momentum in $-z$ direction.
\iteme{= 'FIXT' :} fixed-target experiment, with beam particle
momentum pointing in $+z$ direction.
\iteme{= '3MOM' :} full freedom to specify frame by giving beam
momentum in \ttt{P(1,1)}, \ttt{P(1,2)} and \ttt{P(1,3)} and target
momentum in \ttt{P(2,1)}, \ttt{P(2,2)} and \ttt{P(2,3)} in
common block \ttt{PYJETS}. Particles are assumed on the mass shell,
and energies are calculated accordingly.
\iteme{= '4MOM' :} as \ttt{'3MOM'}, except also energies should be
specified, in \ttt{P(1,4)} and \ttt{P(2,4)}, respectively. The
particles need not be on the mass shell; effective masses are
calculated from energy and momentum. (But note that numerical
precision may suffer; if you know the masses the option \ttt{'5MOM'}
below is preferable.)
\iteme{= '5MOM' :} as \ttt{'3MOM'}, except also energies and masses
should be specified, i.e the full momentum information in
\ttt{P(1,1) - P(1,5)} and \ttt{P(2,1) - P(2,5)} should be given for
beam and target, respectively. Particles need not be on the mass
shell. Space-like virtualities should be stored as $-\sqrt{-m^2}$.
Especially useful for physics with virtual photons. (The virtuality
could be varied from one event to the next, but then it is convenient
to initialize for the lowest virtuality likely to be encountered.)
Four-momentum and mass information must match.
\iteme{= 'USER' :} a run primarily intended to involve Les Houches
Accord external, user-defined processes, see section
\ref{ss:PYnewproc}. Information on incoming beam particles and energies
is read from the \ttt{HEPRUP} common block. In this option, the
\ttt{BEAM}, \ttt{TARGET} and \ttt{WIN} arguments are dummy.
\iteme{= 'NONE' :} there will be no initialization of any
processes, but only of resonance widths and a few other
process-independent variables. Subsequent to such a call,
\ttt{PYEVNT} cannot be used to generate events, so this option
is mainly intended for those who will want to construct their own
events afterwards, but still want to have access to some of the
{\Py} facilities. In this option, the \ttt{BEAM}, \ttt{TARGET} and
\ttt{WIN} arguments are dummy.
\end{subentry}
\iteme{BEAM, TARGET :} character variables to specify beam and
target particles. Upper-case and lower-case letters may be freely
mixed. An antiparticle can be denoted by `bar' at the end of the name
(`$\sim$' is a valid alternative for reasons of backwards compatibility).
It is also possible to leave out the underscore (`\_') directly after `nu'
in neutrino names, and the charge for proton and neutron. The arguments
are dummy when the \ttt{FRAME} argument above is either \ttt{'USER'} or
\ttt{'NONE'}.
\begin{subentry}
\iteme{= 'e-' :} electron.
\iteme{= 'e+' :} positron.
\iteme{= 'nu\_e' :} $\nu_{\e}$.
\iteme{= 'nu\_ebar' :} $\br{\nu}_{\e}$.
\iteme{= 'mu-' :} $\mu^-$.
\iteme{= 'mu+' :} $\mu^+$.
\iteme{= 'nu\_mu' :} $\nu_{\mu}$.
\iteme{= 'nu\_mubar' :} $\br{\nu}_{\mu}$.
\iteme{= 'tau-' :} $\tau^-$.
\iteme{= 'tau+' :} $\tau^+$.
\iteme{= 'nu\_tau' :} $\nu_{\tau}$.
\iteme{= 'nu\_taubar' :} $\br{\nu}_{\tau}$.
\iteme{= 'gamma' :} photon (real, i.e.\ on the mass shell).
\iteme{= 'gamma/e-' :} photon generated by the virtual-photon
flux in an electron beam; \ttt{WIN} below refers to electron,
while photon energy and virtuality varies between events according
to what is allowed by \ttt{CKIN(61) - CKIN(78)}.
\iteme{= 'gamma/e+' :} as above for a positron beam.
\iteme{= 'gamma/mu-' :} as above for a $\mu^-$ beam.
\iteme{= 'gamma/mu+' :} as above for a $\mu^+$ beam.
\iteme{= 'gamma/tau-' :} as above for a $\tau^-$ beam.
\iteme{= 'gamma/tau+' :} as above for a $\tau^+$ beam.
\iteme{= 'pi0' :} $\pi^0$.
\iteme{= 'pi+' :} $\pi^+$.
\iteme{= 'pi-' :} $\pi^-$.
\iteme{= 'n0' :} neutron.
\iteme{= 'nbar0' :} antineutron.
\iteme{= 'p+' :} proton.
\iteme{= 'pbar-' :} antiproton.
\iteme{= 'K+' :} $\K^+$ meson; since parton distributions for strange
hadrons are not available, very simple and untrustworthy recipes are
used for this and subsequent hadrons, see section
\ref{ss:structfun}.
\iteme{= 'K-' :} $\K^-$ meson.
\iteme{= 'KS0' :} $\K_{\mrm{S}}^0$ meson.
\iteme{= 'KL0' :} $\K_{\mrm{L}}^0$ meson.
\iteme{= 'Lambda0' :} $\Lambda$ baryon.
\iteme{= 'Sigma-' :} $\Sigma^-$ baryon.
\iteme{= 'Sigma0' :} $\Sigma^0$ baryon.
\iteme{= 'Sigma+' :} $\Sigma^+$ baryon.
\iteme{= 'Xi-' :} $\Xi^-$ baryon.
\iteme{= 'Xi0' :} $\Xi^0$ baryon.
\iteme{= 'Omega-' :} $\Omega^-$ baryon.
\iteme{= 'pomeron' :} the pomeron $\pomeron$; since pomeron parton
distribution functions have not been defined this option can not
be used currently.
\iteme{= 'reggeon' :} the reggeon $\reggeon$, with comments as for
the pomeron above.
\end{subentry}
\iteme{WIN :} related to energy of system, exact meaning depends on
\ttt{FRAME}.
\begin{subentry}
\iteme{FRAME = 'CMS' :} total energy of system (in GeV).
\iteme{FRAME = 'FIXT' :} momentum of beam particle (in GeV/$c$).
\iteme{FRAME = '3MOM', '4MOM', '5MOM' :} dummy (information is taken from
the \ttt{P} vectors, see above).
\iteme{FRAME = 'USER' :} dummy (information is taken from the \ttt{HEPRUP}
common block, see above).
\iteme{FRAME = 'NONE' :} dummy (no information required).
\end{subentry}
\end{entry}
\drawbox{CALL PYEVNT}\label{p:PYEVNT}
\begin{entry}
\itemc{Purpose:} to generate one event of the type specified by the
\ttt{PYINIT} call, using the traditional `old' underlying-event and
parton-shower model. It is also possible to have \ttt{PYEVNT} call
\ttt{PYEVNW} to access the `new' model that way, see further
\ttt{MSTP(81)}. (This is the main routine, which calls a number
of other routines for specific tasks.)
\end{entry}
\drawbox{CALL PYEVNW}\label{p:PYEVNW}
\begin{entry}
\itemc{Purpose:} to generate one event of the type specified by the
\ttt{PYINIT} call, using the `new' underlying-event and parton-shower
model. (This is the main routine, which calls a number
of other routines for specific tasks.)
\itemc{Warning:} this routine can be used in exactly the same way as
\ttt{PYEVNT}. Technically, one can freely mix calls to the two routines
in the same run, after the \ttt{ PYINIT} call. However, several of the
multiple interactions and shower parameters have a different meaning, or
at least a different proposed best value, which means that caution is
recommended. For instance, a change of $\pTzero$ in \ttt{PARP(82)}
is almost certainly necessary between \ttt{PYEVNT} and \ttt{PYEVNW},
and this require a re-initialization to take effect, so in the end
one cannot mix.
\end{entry}
\drawbox{CALL PYSTAT(MSTAT)}\label{p:PYSTAT}
\begin{entry}
\itemc{Purpose:} to print out cross-sections statistics, decay
widths, branching ratios, status codes and parameter values.
\ttt{PYSTAT} may be called at any time, after the \ttt{PYINIT} call,
e.g.\ at the end of the run, or not at all.
\iteme{MSTAT :} specification of desired information.
\begin{subentry}
\iteme{= 1 :} prints a table of how many events of the different
kinds that have been generated and the corresponding
cross sections. All numbers already include the effects of cuts
required by you in \ttt{PYKCUT}.\\
At the bottom of the listing is also given the total number of warnings
and errors in the current run. (These numbers are reset at each
\ttt{PYINIT} call.) By default only the ten first warnings and errors
are written explicitly; here one may easily see whether many further
occured but were not written in the output. The final number is the
fraction of events that have failed the fragmentation cuts, i.e.\ where,
for one reason or another, the program has had problems fragmenting the
system and has asked for a new hard subprocess.\\
Note that no errors are given on the cross sections. In most cases a
cross section is obtained by Monte Carlo integration during the course
of the run. (Exceptions include e.g. total and elastic hadron--hadron
cross sections, which are parameterized and thus known from the very
onset.) A rule of thumb would then be that the statistical error of a
given subprocess scales like $\delta \sigma / \sigma \approx 1/\sqrt{n}$,
where $n$ is the number of events generated of this kind. In principle,
the numerator of this relation could be decreased by making use of the
full information accumulated during the run, i.e.\ also on the cross
section in those phase space points that are eventually rejected. This
is actually the way the cross section itself is calculated. However, once
you introduce further cuts so that only some fraction of the generated
events survive to the final analysis, you would be back to the simple
$1/\sqrt{n}$ scaling rule for that number of surviving events. Statistical
errors are therefore usually better evaluated within the context of a
specific analysis. Furthermore, systematic errors often dominate over the
statistical ones.\\
Also note that runs with very few events, in addition to having large
errors, tend to have a bias towards overestimating the cross sections.
In a typical case, the average cross section obtained with many runs of
only one event each may be twice that of the correct answer of a single run
with many events. The reason is a `quit while you are ahead' phenomenon,
that an upwards fluctuation in the differential cross section in an early
try gives an acceptable event and thus terminates the run, while a downwards
one leads to rejection and a continuation of the run.
\iteme{= 2 :} prints a table of the resonances defined in the
program, with their particle codes ({\KF}), and all allowed decay
channels. (If the number of generations in \ttt{MSTP(1)} is 3,
however, channels involving fourth-generation particles are not
displayed.) For each decay channel is shown the sequential channel
number (IDC) of the {\Py} decay tables, the decay products (usually
two but sometimes three), the partial decay width,
branching ratio and effective branching ratio (in the event
some channels have been excluded by you).
\iteme{= 3 :} prints a table with the allowed hard interaction
flavours \ttt{KFIN(I,J)} for beam and target particles.
\iteme{= 4 :} prints a table of the kinematical cuts \ttt{CKIN(I)}
set by you in the current run.
\iteme{= 5 :} prints a table with all the values of the status codes
\ttt{MSTP(I)} and the parameters \ttt{PARP(I)} used in the current
run.
\iteme{= 6 :} prints a table of all subprocesses implemented in the
program.
\iteme{= 7 :} prints two tables related to $R$-violating supersymmetry,
where lepton and/or baryon number is not conserved. The first is a
collection of semi-inclusive branching ratios where the entries have
a form like \ttt{\~{}chi\_10 --> nu + q + q}, where a sum has
been performed over all lepton and quark flavours. In the rightmost
column of the table, the number of modes that went into the sum is given.
The purpose of this table is to give a quick overview of the branching
fractions, since there are currently more than 1500 individual
$R$-violating processes
included in the generator. Note that only the pure $1\to 3$ parts of the
3-body modes are included in this sum. If a process can also proceed via
two successive $1\to 2$ branchings (i.e.\ the intermediate resonance is on
shell) the product of these branchings should be added to the number given
in this table. A small list at the bottom of the table shows
the total number of $R$-violating processes in the generator,
the number with non-zero branching ratios in the current run, and the
number with branching ratios larger than $10^{-3}$.
The second table which is printed by this call merely
lists the $R$-violating $\lambda$, $\lambda'$, and $\lambda''$ couplings.
\end{subentry}
\end{entry}
\drawbox{CALL PYFRAM(IFRAME)}\label{p:PYFRAM}
\begin{entry}
\itemc{Purpose:} to transform an event listing between different
reference frames, if so desired. The use of this routine assumes
you do not do any boosts yourself.
\iteme{IFRAME :} specification of frame the event is to be
boosted to.
\begin{subentry}
\iteme{= 1 :} frame specified by you in the \ttt{PYINIT} call.
\iteme{= 2 :} c.m.\ frame of incoming particles.
\iteme{= 3 :} hadronic c.m.\ frame of lepton--hadron interaction
events. Mainly intended for Deeply Inelastic Scattering, but can also
be used in photoproduction. Is not guaranteed to work with the
{\galep} options, however, and so of limited use. Note that both the
lepton and any photons radiated off the lepton remain in the event
listing, and have to be removed separately if you only want to study
the hadronic subsystem.
\end{subentry}
\end{entry}
\drawbox{CALL PYKCUT(MCUT)}\label{p:PYKCUT}
\begin{entry}
\itemc{Purpose:} to enable you to reject a given set of
kinematic variables at an early stage of the generation procedure
(before evaluation of cross sections), so as not to spend unnecessary
time on the generation of events that are not wanted.
The routine will not be called unless you require is by
setting \ttt{MSTP(141) = 1}, and never if `minimum-bias'-type events
(including elastic and diffractive scattering) are to be generated
as well. Furthermore it is never called for user-defined external
processes. A dummy routine \ttt{PYKCUT} is included in the program
file, so as to avoid unresolved external references when the routine
is not used.
\iteme{MCUT :} flag to signal effect of user-defined cuts.
\begin{subentry}
\iteme{= 0 :} event is to be retained and generated in full.
\iteme{= 1 :} event is to be rejected and a new one generated.
\end{subentry}
\itemc{Remark :} at the time of selection, several variables in the
\ttt{MINT} and \ttt{VINT} arrays in the \ttt{PYINT1} common block
contain information that can be used to make the decision. The routine
provided in the program file explicitly reads the variables that
have been defined at the time \ttt{PYKCUT} is called, and also
calculates some derived quantities. The information available
includes subprocess type {\ISUB}, $E_{\mrm{cm}}$, $\hat{s}$, $\hat{t}$,
$\hat{u}$, $\hat{p}_{\perp}$, $x_1$, $x_2$, $x_{\mrm{F}}$, $\tau$, $y$,
$\tau'$, $\cos\hat{\theta}$, and a few more. Some of these may not
be relevant for the process under study, and are then set to zero.
\end{entry}
\subsection{Switches for Event Type and Kinematics Selection}
\label{ss:PYswitchkin}
By default, if {\Py} is run for a hadron collider,
only QCD $2 \to 2$ processes are generated,
composed of hard interactions above $\pTmin=$\ttt{PARP(81)},
with low-$\pT$ processes added on so as to give the full
(parameterized) inelastic, non-diffractive cross section.
In an $\ee$ collider, $\gammaZ$ production is the default, and in
an $\ep$ one it is Deeply Inelastic Scattering. With the help of the
common block \ttt{PYSUBS}, it is possible to select the generation
of another process, or combination of processes. It is also allowed
to restrict the generation to specific incoming partons/particles
at the hard interaction. This often automatically also restricts
final-state flavours but, in processes such as resonance production
or QCD/QED production of new flavours, switches in the {\Py}
program may be used to this end; see section \ref{ss:parapartdat}.
The \ttt{CKIN} array may be used to impose specific kinematics cuts.
You should here be warned that, if kinematical variables are
too strongly restricted, the generation time per event may become
very long. In extreme cases, where the cuts effectively close the
full phase space, the event generation may run into an infinite
loop. The generation of $2 \to 1$ resonance production is performed
in terms of the $\hat{s}$ and $y$ variables, and so the ranges
\ttt{CKIN(1) - CKIN(2)} and \ttt{CKIN(7) - CKIN(8)} may be
arbitrarily restricted without a significant loss of speed.
For $2 \to 2$ processes, $\cos\hat{\theta}$ is added as a third
generation variable, and so additionally the range
\ttt{CKIN(27) - CKIN(28)} may be restricted without any loss of
efficiency.
Effects from initial- and final-state radiation
are not included, since they are not known at the time the
kinematics at the hard interaction is selected. The sharp
kinematical cut-offs that can be imposed on the generation
process are therefore smeared, both by QCD radiation and by
fragmentation. A few examples of such effects follow.
\begin{Itemize}
\item Initial-state radiation implies that each of the two incoming
partons has a non-vanishing $\pT$ when they interact. The hard
scattering subsystem thus receives a net transverse boost,
and is rotated with respect to the beam directions.
In a $2 \to 2$ process, what typically happens is that one of the
scattered partons receives an increased $\pT$, while the $\pT$
of the other parton can be reduced or increased, depending on the
detailed topology.
\item Since the initial-state radiation
machinery assigns space-like virtualities to the incoming partons,
the definitions of $x$ in terms of energy fractions and in terms of
momentum fractions no longer coincide, and so the interacting
subsystem may receive a net longitudinal boost compared with
na\"{\i}ve expectations, as part of the parton-shower machinery.
\item Initial-state radiation gives rise to
additional jets, which in extreme cases may be mistaken for either
of the jets of the hard interaction.
\item Final-state radiation gives rise to additional jets, which
smears the meaning of the basic $2 \to 2$ scattering. The assignment
of soft jets is not unique. The energy of a jet becomes dependent
on the way it is identified, e.g.\ what jet cone size is used.
\item The beam-remnant description assigns primordial $k_{\perp}$
values, which also gives a net $\pT$ shift of the hard-interaction
subsystem; except at low energies this effect is overshadowed by
initial-state radiation, however. Beam remnants may also add
further activity under the `perturbative' event.
\item Fragmentation will further broaden jet profiles, and make
jet assignments and energy determinations even more uncertain.
\end{Itemize}
In a study of events within a given window of
experimentally defined variables, it is up to you to leave
such liberal margins that no events are missed. In other words, cuts
have to be chosen such that a negligible fraction of events migrate
from outside the simulated region to inside the interesting region.
Often this may lead to low efficiency in terms of what fraction of
the generated events are actually of interest to you.
See also section \ref{ss:PYTstarted}.
In addition to the variables found in \ttt{PYSUBS}, also those in the
\ttt{PYPARS} common block may be used to select exactly what one wants
to have simulated. These possibilities will be described in the
following section.
The notation used above and in the following is that `$\hat{~}$'
denotes internal variables in the hard-scattering subsystem,
while `$^*$' is for variables in the c.m.\ frame of the
event as a whole.
\drawbox{COMMON/PYSUBS/MSEL,MSELPD,MSUB(500),KFIN(2,-40:40),CKIN(200)}%
\label{p:PYSUBS}
\begin{entry}
\itemc{Purpose:} to allow you to run the program with any desired
subset of processes, or restrict flavours or kinematics. If the
default values, denoted below by (D = \ldots), are not satisfactory,
they must be changed before the \ttt{PYINIT} call.
\boxsep
\iteme{MSEL :}\label{p:MSEL} (D = 1) a switch to select between full
user control and some preprogrammed alternatives.
\begin{subentry}
\iteme{= 0 :} desired subprocesses have to be switched on in
\ttt{MSUB}, i.e.\ full user control.
\iteme{= 1 :} depending on incoming particles, different
alternatives are used. \\
Lepton--lepton: $\Z$ or $\W$ production ({\ISUB} = 1 or 2). \\
Lepton--hadron: Deeply Inelastic Scattering ({\ISUB} = 10; this option
is now out of date for most applications, superseded by the
{\galep} machinery). \\
Hadron--hadron: QCD high-$\pT$ processes ({\ISUB} = 11, 12, 13, 28,
53, 68); additionally low-$\pT$ production if
\ttt{CKIN(3)} $<$ \ttt{PARP(81)} or \ttt{PARP(82)}, depending on
\ttt{MSTP(82)} ({\ISUB} = 95). If low-$\pT$ is switched on, the other
\ttt{CKIN} cuts are not used. \\
A resolved photon counts as hadron. When the photon is not resolved,
the following cases are possible. \\
Photon--lepton: Compton scattering ({\ISUB} = 34). \\
Photon--hadron: photon-parton scattering ({\ISUB} = 33, 34, 54). \\
Photon--photon: fermion pair production ({\ISUB} = 58).\\
When photons are given by the {\galep} argument in the
\ttt{PYINIT} call, the outcome depends on the \ttt{MSTP(14)} value.
Default is a mixture of many kinds of processes, as described
in section \ref{ss:photoanddisclass}.
\iteme{= 2 :} as \ttt{MSEL = 1} for lepton--lepton, lepton--hadron
and unresolved photons. For hadron--hadron (including resolved
photons) all QCD processes, including low-$\pT$, single and
double diffractive and elastic scattering, are included ({\ISUB} =
11, 12, 13, 28, 53, 68, 91, 92, 93, 94, 95). The \ttt{CKIN} cuts are
here not used.\\
For photons given with the {\galep} argument in the
\ttt{PYINIT} call, the above processes are replaced by other ones
that also include the photon virtuality in the cross sections. The
principle remains to include both high- and low-$\pT$ processes,
however.
\iteme{= 4 :} charm ($\c\cbar$) production with massive matrix
elements ({\ISUB} = 81, 82, 84, 85).
\iteme{= 5 :} bottom ($\b\bbar$) production with massive matrix
elements ({\ISUB} = 81, 82, 84, 85).
\iteme{= 6 :} top ($\t\tbar$) production with massive matrix
elements ({\ISUB} = 81, 82, 84, 85).
\iteme{= 7 :} fourth generation $\b'$ ($\b'\bbar'$) production with
massive matrix elements ({\ISUB} = 81, 82, 84, 85).
\iteme{= 8 :} fourth generation $\t'$ ($\t'\tbar'$) production with
massive matrix elements ({\ISUB} = 81, 82, 84, 85).
\iteme{= 10 :} prompt photons ({\ISUB} = 14, 18, 29).
\iteme{= 11 :} $\Z^0$ production ({\ISUB} = 1).
\iteme{= 12 :} $\W^{\pm}$ production ({\ISUB} = 2).
\iteme{= 13 :} $\Z^0$ + jet production ({\ISUB} = 15, 30).
\iteme{= 14 :} $\W^{\pm}$ + jet production ({\ISUB} = 16, 31).
\iteme{= 15 :} pair production of different combinations of $\gamma$,
$\Z^0$ and $\W^{\pm}$ (except $\gamma\gamma$; see \ttt{MSEL} = 10)
({\ISUB} = 19, 20, 22, 23, 25).
\iteme{= 16 :} $\hrm^0$ production ({\ISUB} = 3, 102, 103, 123, 124).
\iteme{= 17 :} $\hrm^0 \Z^0$ or $\hrm^0 \W^{\pm}$ ({\ISUB} = 24, 26).
\iteme{= 18 :} $\hrm^0$ production, combination relevant for $\ee$
annihilation ({\ISUB} = 24, 103, 123, 124).
\iteme{= 19 :} $\hrm^0$, $\H^0$ and $\A^0$ production, excepting pair
production ({\ISUB} = 24, 103, 123, 124, 153, 158, 171, 173, 174, 176,
178, 179).
\iteme{= 21 :} $\Z'^0$ production ({\ISUB} = 141).
\iteme{= 22 :} $\W'^{\pm}$ production ({\ISUB} = 142).
\iteme{= 23 :} $\H^{\pm}$ production ({\ISUB} = 143).
\iteme{= 24 :} $\R^0$ production ({\ISUB} = 144).
\iteme{= 25 :} $\L_{\Q}$ (leptoquark) production ({\ISUB} = 145, 162,
163, 164).
\iteme{= 35 :} single bottom production by $\W$ exchange ({\ISUB} = 83).
\iteme{= 36 :} single top production by $\W$ exchange ({\ISUB} = 83).
\iteme{= 37 :} single $\b'$ production by $\W$ exchange ({\ISUB} = 83).
\iteme{= 38 :} single $\t'$ production by $\W$ exchange ({\ISUB} = 83).
\iteme{= 39 :} all MSSM processes except Higgs production.
\iteme{= 40 :} squark and gluino production ({\ISUB} = 243, 244, 258, 259,
271--280).
\iteme{= 41 :} stop pair production ({\ISUB} = 261--265).
\iteme{= 42 :} slepton pair production ({\ISUB} = 201--214).
\iteme{= 43 :} squark or gluino with chargino or neutralino,
({\ISUB} = 237--242, 246--256).
\iteme{= 44 :} chargino--neutralino pair production ({\ISUB} = 216--236).
\iteme{= 45 :} sbottom production ({\ISUB} = 281--296).
\iteme{= 50 :} pair production of technipions and gauge bosons by
$\pi^{0,\pm}_{\mrm{tc}}/\omega^0_{\mrm{tc}}$ exchange
({\ISUB} = 361--377).
\iteme{= 51 :} standard QCD $2 \to 2$ processes 381--386, with
possibility to introduce compositeness/technicolor modifications,
see \ttt{ITCM(5)}.
\iteme{= 61 :} charmonimum production in the NRQCD framework,
({\ISUB} = 421--439).
\iteme{= 62 :} bottomonimum production in the NRQCD framework,
({\ISUB} = 461--479).
\iteme{= 63 :} both charmonimum and bottomonimum production in the \
NRQCD framework, ({\ISUB} = 421--439, 461--479).
\end{subentry}
\boxsep
\iteme{MSUB :}\label{p:MSUB} (D = 500*0) array to be set when
\ttt{MSEL = 0} (for \ttt{MSEL} $\geq 1$ relevant entries are set in
\ttt{PYINIT}) to choose which subset of subprocesses to include
in the generation. The ordering follows the {\ISUB} code given in
section \ref{ss:ISUBcode} (with comments as given there).
\begin{subentry}
\iteme{MSUB(ISUB) = 0 :} the subprocess is excluded.
\iteme{MSUB(ISUB) = 1 :} the subprocess is included.
\itemc{Note:} when \ttt{MSEL = 0}, the \ttt{MSUB} values set by
you are never changed by {\Py}. If you want to combine several
different `subruns', each with its own \ttt{PYINIT} call, into one
single run, it is up to you to remember not only to switch on
the new processes before each new \ttt{PYINIT} call, but also to
switch off the old ones that are no longer desired.
\end{subentry}
\boxsep
\iteme{KFIN(I,J) :}\label{p:KFIN} provides an option to selectively
switch on and
off contributions to the cross sections from the different incoming
partons/particles at the hard interaction. In combination with the
{\Py} resonance decay switches, this also allows you to set
restrictions on flavours appearing in the final state.
\begin{subentry}
\iteme{I :} is 1 for beam side of event and 2 for target side.
\iteme{J :} enumerates flavours according to the {\KF} code; see section
\ref{ss:codes}.
\iteme{KFIN(I,J) = 0 :} the parton/particle is forbidden.
\iteme{KFIN(I,J) = 1 :} the parton/particle is allowed.
\itemc{Note:} by default, the following are switched on: $\d$, $\u$,
$\s$, $\c$, $\b$, $\e^-$, $\nu_{\e}$, $\mu^-$, $\nu_{\mu}$, $\tau^-$,
$\nu_{\tau}$, $\g$, $\gamma$, $\Z^0$, $\W^+$ and their antiparticles.
In particular, top is off, and has to be switched on explicitly if
needed.
\end{subentry}
\boxsep
\iteme{CKIN :}\label{p:CKIN} kinematics cuts that can be set by you
before the \ttt{PYINIT} call, and that affect the region of phase space
within which events are generated. Some cuts are `hardwired' while
most are `softwired'. The hardwired ones are directly related to the
kinematical variables used in the event selection procedure,
and therefore have negligible effects on program efficiency.
The most important of these are \ttt{CKIN(1) - CKIN(8)},
\ttt{CKIN(27) - CKIN(28)}, and \ttt{CKIN(31) - CKIN(32)}.
The softwired ones are most of the remaining ones, that cannot
be fully taken into account
in the kinematical variable selection, so that generation in
constrained regions of phase space may be slow. In extreme
cases the phase space may be so small that the maximization
procedure fails to find any allowed points at all (although some
small region might still exist somewhere), and therefore switches
off some subprocesses, or aborts altogether.
\iteme{CKIN(1), CKIN(2) :} (D = 2., $-1.$ GeV) range of allowed
$\hat{m} = \sqrt{\hat{s}}$ values. If \ttt{CKIN(2)} $< 0.$, the upper
limit is inactive.
\iteme{CKIN(3), CKIN(4) :} (D = 0., $-1.$ GeV) range of allowed
$\hat{p}_{\perp}$ values for hard $2 \to 2$ processes, with
transverse momentum $\hat{p}_{\perp}$ defined in the rest frame of
the hard interaction. If \ttt{CKIN(4)} $< 0.$, the upper limit is
inactive. For processes that are singular in the limit
$\hat{p}_{\perp} \to 0$
(see \ttt{CKIN(6)}), \ttt{CKIN(5)} provides an additional constraint.
The \ttt{CKIN(3)} and \ttt{CKIN(4)} limits can also be used in
$2 \to 1 \to 2$ processes. Here, however, the product
masses are not known and hence are assumed to be vanishing in the event
selection. The actual $\pT$ range for massive products is thus
shifted downwards with respect to the nominal one.
\begin{subentry}
\itemc{Note 1:} for processes that are singular in the limit
$\hat{p}_{\perp} \to 0$, a careful choice of \ttt{CKIN(3)} value is
not only a matter of technical convenience, but a requirement for
obtaining sensible results. One example is the hadroproduction of
a $\W^{\pm}$ or $\Z^0$ gauge boson together with a jet, discussed
in section \ref{sss:WZclass}. Here the point is that this is a
first-order process (in $\alphas$), correcting the zeroth-order
process of a $\W^{\pm}$ or $\Z^0$ without any jet. A full first-order
description would also have to include virtual corrections in the
low-$\hat{p}_{\perp}$ region.\\
Generalizing also to other processes, the simple-minded higher-order
description breaks down when \ttt{CKIN(3)} is selected so small that
the higher-order process cross section corresponds to a non-negligible
fraction of the lower-order one. This number will vary depending on
the process considered and the c.m. energy used, but could easily be
tens of GeV rather than the default 1 GeV provided as technical
cut-off in \ttt{CKIN(5)}. Processes singular in $\hat{p}_{\perp} \to 0$
should therefore only be used to describe the high-$\pT$ behaviour,
while the lowest-order process complemented with parton showers
should give the inclusive distribution and in particular the one
at small $\pT$ values.\\
Technically the case of QCD production of two jets is slightly more
complicated, and involves eikonalization to multiple parton--parton
scattering, section \ref{ss:multint}, but again the conclusion
is that the processes have to be handled with care at small $\pT$
values.
\itemc{Note 2:} there are a few situations in which \ttt{CKIN(3)}
may be overwritten; especially when different subprocess classes
are mixed in $\gamma\p$ or $\gamma\gamma$ collisions, see section
\ref{ss:photoanddisclass}.
\end{subentry}
\iteme{CKIN(5) :} (D = 1. GeV) lower cut-off on $\hat{p}_{\perp}$ values,
in addition to the \ttt{CKIN(3)} cut above, for processes that are
singular in the limit $\hat{p}_{\perp} \to 0$ (see \ttt{CKIN(6)}).
\iteme{CKIN(6) :} (D = 1. GeV) hard $2 \to 2$ processes, which do not
proceed only via an intermediate resonance (i.e.\ are $2 \to 1 \to 2$
processes), are classified as singular in the limit
$\hat{p}_{\perp} \to 0$ if either or both of the two final-state
products has a mass $m <$ \ttt{CKIN(6)}.
\iteme{CKIN(7), CKIN(8) :} (D = $-10.$, 10.) range of allowed scattering
subsystem rapidities $y = y^*$ in the c.m.\ frame of the event,
where $y = (1/2) \ln(x_1/x_2)$. (Following the notation of this
section, the variable should be given as $y^*$, but because of its
frequent use, it was called $y$ in section \ref{ss:kinemtwo}.)
\iteme{CKIN(9), CKIN(10) :} (D = $-40.$, 40.) range of allowed (true)
rapidities for the product with largest rapidity in a $2 \to 2$ or a
$2 \to 1 \to 2$ process, defined in the c.m.\ frame of the event,
i.e.\ $\max(y^*_3, y^*_4)$. Note that rapidities are counted with sign,
i.e.\ if $y^*_3 = 1$ and $y^*_4 = -2$ then $\max(y^*_3, y^*_4) = 1$.
\iteme{CKIN(11), CKIN(12) :} (D = $-40.$, 40.) range of allowed (true)
rapidities for the product with smallest rapidity in a $2 \to 2$ or a
$2 \to 1 \to 2$ process, defined in the c.m.\ frame of the event,
i.e.\ $\min(y^*_3, y^*_4)$. Consistency thus requires
\ttt{CKIN(11)} $\leq$ \ttt{CKIN(9)} and
\ttt{CKIN(12)} $\leq$ \ttt{CKIN(10)}.
\iteme{CKIN(13), CKIN(14) :} (D = $-40.$, 40.) range of allowed
pseudorapidities for the product with largest pseudorapidity
in a $2 \to 2$ or a $2 \to 1 \to 2$ process, defined in the c.m.\
frame of the event, i.e.\ $\max(\eta^*_3, \eta^*_4)$. Note that
pseudorapidities are counted with sign, i.e.\ if $\eta^*_3 = 1$ and
$\eta^*_4 = -2$ then $\max(\eta^*_3, \eta^*_4) = 1$.
\iteme{CKIN(15), CKIN(16) :} (D = $-40.$, 40.) range of allowed
pseudorapidities for the product with smallest pseudorapidity
in a $2 \to 2$ or a $2 \to 1 \to 2$ process, defined in the c.m.\
frame of the event, i.e.\ $\min(\eta^*_3, \eta^*_4)$. Consistency
thus requires \ttt{CKIN(15)} $\leq$ \ttt{CKIN(13)} and
\ttt{CKIN(16)} $\leq$ \ttt{CKIN(14)}.
\iteme{CKIN(17), CKIN(18) :} (D = $-1.$, 1.) range of allowed
$\cos\theta^*$ values for the product with largest $\cos\theta^*$
value in a $2 \to 2$ or a $2 \to 1 \to 2$ process, defined in the
c.m.\ frame of the event, i.e.\ $\max(\cos\theta^*_3,\cos\theta^*_4)$.
\iteme{CKIN(19), CKIN(20) :} (D = $-1.$, 1.) range of allowed
$\cos\theta^*$ values for the product with smallest $\cos\theta^*$
value in a $2 \to 2$ or a $2 \to 1 \to 2$ process, defined in the
c.m.\ frame of the event, i.e.\ $\min(\cos\theta^*_3,\cos\theta^*_4)$.
Consistency thus requires \ttt{CKIN(19)} $\leq$ \ttt{CKIN(17)} and
\ttt{CKIN(20)} $\leq$ \ttt{CKIN(18)}.
\iteme{CKIN(21), CKIN(22) :} (D = 0., 1.) range of allowed $x_1$ values
for the parton on side 1 that enters the hard interaction.
\iteme{CKIN(23), CKIN(24) :} (D = 0., 1.) range of allowed $x_2$ values
for the parton on side 2 that enters the hard interaction.
\iteme{CKIN(25), CKIN(26) :} (D = $-1.$, 1.) range of allowed Feynman-$x$
values, where $x_{\mrm{F}} = x_1 - x_2$.
\iteme{CKIN(27), CKIN(28) :} (D = $-1.$, 1.) range of allowed
$\cos\hat{\theta}$ values in a hard $2 \to 2$ scattering, where
$\hat{\theta}$ is the scattering angle in the rest frame of the
hard interaction.
\iteme{CKIN(31), CKIN(32) :} (D = 2., $-1.$ GeV) range of allowed
$\hat{m}' = \sqrt{\hat{s}'}$ values, where $\hat{m}'$ is the mass of
the complete three- or four-body final state in $2 \to 3$ or
$2 \to 4$ processes (while $\hat{m}$, constrained in \ttt{CKIN(1)}
and \ttt{CKIN(2)}, here corresponds to the one- or two-body central
system). If \ttt{CKIN(32)} $< 0.$, the upper limit is inactive.
\iteme{CKIN(35), CKIN(36) :} (D = 0., $-1.$ GeV$^2$) range of allowed
$|\hat{t}| = - \hat{t}$ values in $2 \to 2$ processes. Note that
for Deeply Inelastic Scattering this is nothing but the $Q^2$ scale,
in the limit that initial- and final-state radiation is neglected.
If \ttt{CKIN(36)} $< 0.$, the upper limit is inactive.
\iteme{CKIN(37), CKIN(38) :} (D = 0., $-1.$ GeV$^2$) range of allowed
$|\hat{u}| = - \hat{u}$ values in $2 \to 2$ processes. If
\ttt{CKIN(38)} $< 0.$, the upper limit is inactive.
\iteme{CKIN(39), CKIN(40) :} (D = 4., $-1.$ GeV$^2$) the $W^2$ range
allowed in DIS processes, i.e.\ subprocess number 10. If
\ttt{CKIN(40)} $< 0.$, the upper limit is inactive. Here $W^2$ is
defined in terms of $W^2 = Q^2 (1-x)/x$. This formula is not quite
correct, in that \textit{(i)} it neglects the target mass (for a
proton), and \textit{(ii)} it neglects initial-state photon radiation
off the incoming electron. It should be good enough for loose cuts,
however. These cuts are not checked if process 10 is called for two
lepton beams.
\iteme{CKIN(41) - CKIN(44) :} (D = 12., $-1.$, 12., $-1.$ GeV) range of
allowed mass values of the two (or one) resonances produced in a `true'
$2 \to 2$ process, i.e.\ one not (only) proceeding through a single
$s$-channel resonance ($2 \to 1 \to 2$). (These are the ones listed
as $2 \to 2$ in the tables in section \ref{ss:ISUBcode}.)
Only particles with a width above \ttt{PARP(41)} are considered as
bona fide resonances and tested against the \ttt{CKIN}
limits; particles with a smaller width are put on the mass shell
without applying any cuts. The exact interpretation of the \ttt{CKIN}
variables depends on the flavours of the two produced resonances. \\
For two resonances like $\Z^0 \W^+$ (produced from
$\f_i \fbar_j \to \Z^0\W^+$), which are not identical and which are not
each other's antiparticles, one has \\
\ttt{CKIN(41)} $< m_1 <$ \ttt{CKIN(42)}, and \\
\ttt{CKIN(43)} $< m_2 <$ \ttt{CKIN(44)}, \\
where $m_1$ and $m_2$ are the actually generated masses of the two
resonances, and 1 and 2 are defined by the order in which they are
given in the production process specification. \\
For two resonances like $\Z^0 \Z^0$, which are identical, or
$\W^+ \W^-$, which are each other's antiparticles, one instead
has \\
\ttt{CKIN(41)} $< \min(m_1,m_2) <$ \ttt{CKIN(42)}, and \\
\ttt{CKIN(43)} $< \max(m_1,m_2) <$ \ttt{CKIN(44)}. \\
In addition, whatever limits are set on \ttt{CKIN(1)} and, in
particular, on \ttt{CKIN(2)} obviously affect the masses actually
selected.
\begin{subentry}
\itemc{Note 1:} if \ttt{MSTP(42) = 0}, so that no mass smearing is
allowed, the \ttt{CKIN} values have no effect (the same as for
particles with too narrow a width).
\itemc{Note 2:} if \ttt{CKIN(42)} $<$ \ttt{CKIN(41)} it means that
the \ttt{CKIN(42)} limit is inactive; correspondingly, if
\ttt{CKIN(44)} $< $\ttt{CKIN(43)} then \ttt{CKIN(44)} is inactive.
\itemc{Note 3:} if limits are active and the resonances are
identical, it is up to you to ensure that
\ttt{CKIN(41)} $\leq$ \ttt{CKIN(43)} and
\ttt{CKIN(42)} $\leq$ \ttt{CKIN(44)}.
\itemc{Note 4:} for identical resonances, it is not possible to
preselect which of the resonances is the lighter one; if, for
instance, one
$\Z^0$ is to decay to leptons and the other to quarks, there is no
mechanism to guarantee that the lepton pair has a mass smaller
than the quark one.
\itemc{Note 5:} the \ttt{CKIN} values are applied to all relevant
$2 \to 2$ processes equally, which may not be what one desires if
several processes are generated simultaneously. Some caution is
therefore urged in the use of the \ttt{CKIN(41) - CKIN(44)} values.
Also in other respects, you are recommended to take proper
care: if a $\Z^0$ is only allowed to decay into $\b\bbar$, for example,
setting its mass range to be 2--8 GeV is obviously not a
good idea.
\end{subentry}
\iteme{CKIN(45) - CKIN(48) :} (D = 12., $-1.$, 12., $-1.$ GeV) range of
allowed mass values of the two (or one) secondary resonances produced in
a $2 \to 1 \to 2$ process (like $\g \g \to \hrm^0 \to \Z^0 \Z^0$) or
even a $2 \to 2 \to 4$ (or 3) process (like
$\q \qbar \to \Z^0 \hrm^0 \to \Z^0 \W^+ \W^-$). Note that these
\ttt{CKIN} values only affect the secondary resonances; the
primary ones are constrained by \ttt{CKIN(1)}, \ttt{CKIN(2)} and
\ttt{CKIN(41) - CKIN(44)} (indirectly, of course, the choice of
primary resonance masses affects the allowed mass range for the
secondary ones). What is considered to be a resonance is defined
by \ttt{PARP(41)}; particles with a width smaller than this are
automatically put on the mass shell. The description closely
parallels the one given for \ttt{CKIN(41) - CKIN(44)}. Thus, for
two resonances that are not identical or each other's
antiparticles, one has \\
\ttt{CKIN(45)} $< m_1 <$ \ttt{CKIN(46)}, and \\
\ttt{CKIN(47)} $< m_2 <$ \ttt{CKIN(48)}, \\
where $m_1$ and $m_2$ are the actually generated masses of the two
resonances, and 1 and 2 are defined by the order in which they are
given in the decay channel specification in the program (see e.g.
output from \ttt{PYSTAT(2)} or \ttt{PYLIST(12)}). For two
resonances that are identical or each other's antiparticles,
one instead has \\
\ttt{CKIN(45)} $< \min(m_1,m_2) <$ \ttt{CKIN(46)}, and \\
\ttt{CKIN(47)} $< \max(m_1,m_2) <$ \ttt{CKIN(48)}.
\begin{subentry}
\itemc{Notes 1 - 5:} as for \ttt{CKIN(41) - CKIN(44)}, with trivial
modifications.
\itemc{Note 6:} setting limits on secondary resonance masses is
possible in any of the channels of the allowed types (see above).
However, so far only $\hrm^0 \to \Z^0 \Z^0$ and $\hrm^0 \to \W^+ \W^-$
have been fully implemented, such that an arbitrary mass range
below the na\"{\i}ve mass threshold may be picked. For other possible
resonances, any restrictions made on the allowed mass range
are not reflected in the cross section; and further it is not
recommendable to pick mass windows that make a decay on the
mass shell impossible.
\end{subentry}
\iteme{CKIN(49) - CKIN(50) :} allow minimum mass limits to be passed
from \ttt{PYRESD} to \ttt{PYOFSH}. They are used for tertiary and higher
resonances, i.e.\ those not controlled by \ttt{CKIN(41) - CKIN(48)}.
They should not be touched by the user.
\iteme{CKIN(51) - CKIN(56) :} (D = 0., $-1.$, 0., $-1.$, 0., $-1.$ GeV)
range of allowed transverse momenta in a true $2 \to 3$ process. This
means subprocesses such as 121--124 for $\hrm^0$ production, and
their $\H^0$, $\A^0$ and $\H^{\pm\pm}$ equivalents.
\ttt{CKIN(51) - CKIN(54)} corresponds
to $\pT$ ranges for scattered partons, in order of appearance,
i.e.\ \ttt{CKIN(51) - CKIN(52)} for the parton scattered off the beam
and \ttt{CKIN(53) - CKIN(54)} for the one scattered off the target.
\ttt{CKIN(55)} and \ttt{CKIN(56)} here sets $\pT$ limits for the
third product, the $\hrm^0$, i.e.\ the \ttt{CKIN(3)} and \ttt{CKIN(4)}
values have no effect for this process. Since the $\pT$ of the Higgs
is not one of the primary variables selected, any constraints here
may mean reduced Monte Carlo efficiency, while for these processes
\ttt{CKIN(51) - CKIN(54)} are `hardwired' and therefore do not cost
anything. As usual, a negative value implies that the upper limit is
inactive.
\iteme{CKIN(61) - CKIN(78) :} allows to restrict the range of kinematics
for the photons generated off the lepton beams with the
{\galep} option of \ttt{PYINIT}. In each quartet of numbers,
the first two corresponds to the range allowed on incoming side 1 (beam)
and the last two to side 2 (target). The cuts are only applicable for a
lepton beam. Note that the $x$ and $Q^2$ ($P^2$) variables are the basis
for the generation, and so can be restricted with no loss of
efficiency. For leptoproduction (i.e.\ lepton on hadron) the $W$ is
uniquely given by the one $x$ value of the problem, so here also $W$ cuts
are fully efficient. The other cuts may imply a slowdown of the program,
but not as much as if equivalent cuts only are introduced after events
are fully generated. See \cite{Fri00} for details.
\iteme{CKIN(61) - CKIN(64) :} (D = 0.0001, 0.99, 0.0001, 0.99) allowed
range for the energy fractions $x$ that the photon take of the respective
incoming lepton energy. These fractions are defined in the
c.m.\ frame of the collision, and differ from energy fractions
as defined in another frame. (Watch out at HERA!) In order to
avoid some technical problems, absolute lower and upper limits
are set internally at 0.0001 and 0.9999.
\iteme{CKIN(65) - CKIN(68) :} (D = 0., $-1.$, 0., $-1.$ GeV$^2$) allowed
range for the space-like virtuality of the photon, conventionally called
either $Q^2$ or $P^2$, depending on process. A negative number means that
the upper limit is inactive, i.e.\ purely given by kinematics. A nonzero
lower limit is implicitly given by kinematics constraints.
\iteme{CKIN(69) - CKIN(72) :} (D = 0., $-1.$, 0., $-1.$) allowed range of
the scattering angle $\theta$ of the lepton, defined in the c.m.\ frame
of the event. (Watch out at HERA!) A negative number means that
the upper limit is inactive, i.e.\ equal to $\pi$.
\iteme{CKIN(73) - CKIN(76) :} (D = 0.0001, 0.99, 0.0001, 0.99) allowed
range for the light-cone fraction $y$ that the photon take of the \
respective incoming lepton energy. The light-cone is defined by the
four-momentum of the lepton or hadron on the other side of the
event (and thus deviates from true light-cone fraction by mass
effects that normally are negligible). The $y$ value is related to
the $x$ and $Q^2$ ($P^2$) values by $y = x + Q^2/s$ if mass terms are
neglected.
\iteme{CKIN(77), CKIN(78) :} (D = 2., $-1.$ GeV) allowed range for $W$,
i.e.\ either the photon--hadron or photon--photon invariant mass. A
negative number means that the upper limit is inactive.
\end{entry}
\subsection{The General Switches and Parameters}
\label{ss:PYswitchpar}
The \ttt{PYPARS} common block contains the status code and parameters
that regulate the performance of the program. All of them are
provided with sensible default values, so that a novice user can
neglect them, and only gradually explore the full range of
possibilities. Some of the switches and parameters in \ttt{PYPARS}
will be described later, in the shower and beam-remnants sections.
\drawbox{COMMON/PYPARS/MSTP(200),PARP(200),MSTI(200),PARI(200)}%
\label{p:PYPARS1}
\begin{entry}
\itemc{Purpose:} to give access to status code and parameters that
regulate the performance of the program. If the default values,
denoted below by (D = \ldots), are not satisfactory, they must in
general be changed before the \ttt{PYINIT} call. Exceptions, i.e.\
variables that can be changed for each new event, are denoted by
(C).
\iteme{MSTP(1) :}\label{p:MSTP} (D = 3) maximum number of generations.
Automatically set $\leq 4$.
\iteme{MSTP(2) :} (D = 1) calculation of $\alphas$ at hard interaction,
in the routine \ttt{PYALPS}.
\begin{subentry}
\iteme{= 0 :} $\alphas$ is fixed at value \ttt{PARU(111)}.
\iteme{= 1 :} first-order running $\alphas$.
\iteme{= 2 :} second-order running $\alphas$.
\end{subentry}
\iteme{MSTP(3) :} (D = 2) selection of $\Lambda$ value in $\alphas$
for \ttt{MSTP(2)} $\geq 1$.
\begin{subentry}
\iteme{= 1 :} $\Lambda$ is given by \ttt{PARP(1)} for hard
interactions, by \ttt{PARP(61)} for space-like showers,
by \ttt{PARP(72)} for time-like showers not from a resonance decay,
and by \ttt{PARJ(81)} for time-like ones from a resonance decay
(including e.g.\ $\gamma/\Z^0 \to \q\qbar$ decays, i.e.\ conventional
$\ee$ physics). This $\Lambda$ is assumed
to be valid for 5 flavours; for the hard interaction the number of
flavours assumed can be changed by \ttt{MSTU(112)}.
\iteme{= 2 :} $\Lambda$ value is chosen according to the
parton-distribution-function para\-meteri\-zations. The choice is
always based on the proton parton-distribution set selected, i.e.\
is unaffected by pion and photon parton-distribution selection.
All the $\Lambda$ values
are assumed to refer to 4 flavours, and \ttt{MSTU(112)} is set
accordingly. This $\Lambda$ value is used both for the hard
scattering and the initial- and final-state radiation. The
ambiguity in the choice of the $Q^2$ argument still remains (see
\ttt{MSTP(32)}, \ttt{MSTP(64)} and \ttt{MSTJ(44)}). This $\Lambda$
value is used also for \ttt{MSTP(57) = 0}, but the sensible choice
here would be to use \ttt{MSTP(2) = 0} and have no initial- or
final-state radiation. This option does \textit{not} change the
\ttt{PARJ(81)} value of time-like parton showers in resonance decays,
so that LEP experience on this specific parameter is not overwritten
unwittingly. Therefore \ttt{PARJ(81)} can be updated completely
independently.
\iteme{= 3 :} as \ttt{= 2}, except that here also \ttt{PARJ(81)} is
overwritten in accordance with the $\Lambda$ value of the proton
parton-distribution-function set.
\end{subentry}
\iteme{MSTP(4) :} (D = 0) treatment of the Higgs sector,
predominantly the neutral one.
\begin{subentry}
\iteme{= 0 :} the $\hrm^0$ is given the Standard Model Higgs
couplings, while $\H^0$ and $\A^0$ couplings should be set by
you in \ttt{PARU(171) - PARU(175)} and
\ttt{PARU(181) - PARU(185)}, respectively.
\iteme{= 1 :} you should set couplings for all three Higgs bosons,
for the $\hrm^0$ in \ttt{PARU(161) - PARU(165)}, and for the
$\H^0$ and $\A^0$ as above.
\iteme{= 2 :} the mass of $\hrm^0$ in \ttt{PMAS(25,1)} and the
$\tan\beta$ value in \ttt{PARU(141)} are used to derive
$\H^0$, $\A^0$ and $\H^{\pm}$ masses, and $\hrm^0$, $\H^0$,
$\A^0$ and $\H^{\pm}$ couplings, using the relations of the Minimal
Supersymmetric extension of the Standard Model at Born level
\cite{Gun90}. Existing masses and couplings are overwritten by the
derived values. See section \ref{sss:extneutHclass} for discussion
on parameter constraints.
\iteme{= 3:} as \ttt{= 2}, but using relations at the one-loop level.
This option is not yet implemented as such. However, if you initialize
the SUSY machinery with \ttt{IMSS(1) = 1}, then the SUSY
parameters will be used to calculate also Higgs masses and couplings.
These are stored in the appropriate slots, and the value of \ttt{MSTP(4)}
is overwritten to 1.
\end{subentry}
\iteme{MSTP(7) :} (D = 0) choice of heavy flavour in subprocesses
81--85. Does not apply for \ttt{MSEL = 4 - 8}, where the \ttt{MSEL} value
always takes precedence.
\begin{subentry}
\iteme{= 0 :} for processes 81--84 (85) the `heaviest' flavour
allowed for gluon (photon) splitting into a quark--antiquark
(fermion--antifermion) pair, as set in the \ttt{MDME} array.
Note that `heavy' is defined as the one with largest {\KF} code,
so that leptons take precedence if they are allowed.
\iteme{= 1 - 8 :} pick this particular quark flavour; e.g.,
\ttt{MSTP(7) = 6} means that top will be produced.
\iteme{= 11 - 18 :} pick this particular lepton flavour. Note that
neutrinos are not possible, i.e.\ only 11, 13, 15 and 17 are
meaningful alternatives. Lepton pair production can only occur in
process 85, so if any of the other processes have been switched on
they are generated with the same flavour as would be obtained in
the option \ttt{MSTP(7) = 0}.
\end{subentry}
\iteme{MSTP(8) :} (D = 0) choice of electroweak parameters to use
in the decay widths of resonances ($\W$, $\Z$, $\hrm$, \ldots) and
cross sections (production of $\W$'s, $\Z$'s, $\hrm$'s, \ldots).
\begin{subentry}
\iteme{= 0 :} everything is expressed in terms of a running
$\alphaem(Q^2)$ and a fixed $\ssintw$, i.e.\ $G_{\F}$ is nowhere
used.
\iteme{= 1 :} a replacement is made according to
$\alphaem(Q^2) \to \sqrt{2} G_{\F} m_{\W}^2 \ssintw / \pi$
in all widths and cross sections. If $G_{\F}$ and $m_{\Z}$ are
considered as given, this means that $\ssintw$ and $m_{\W}$ are the
only free electroweak parameter.
\iteme{= 2 :} a replacement is made as for \ttt{= 1}, but additionally
$\ssintw$ is constrained by the relation
$\ssintw = 1 - m_{\W}^2/m_{\Z}^2$.
This means that $m_{\W}$ remains as a free parameter, but that the
$\ssintw$ value in \ttt{PARU(102)} is never used, \textit{except} in
the vector couplings in the combination $v = a - 4 \ssintw e$.
This latter degree of freedom enters e.g.\ for forward-backward
asymmetries in $\Z^0$ decays.
\itemc{Note:} this option does not affect the emission of real photons
in the initial and final state, where $\alphaem$ is always used. However,
it does affect also purely electromagnetic hard processes, such as
$\q \qbar \to \gamma \gamma$.
\end{subentry}
\iteme{MSTP(9) :} (D = 0) inclusion of top (and fourth generation) as
allowed remnant flavour $\q'$ in processes that involve $\q \to \q' + \W$
branchings as part of the overall process, and where the matrix elements
have been calculated under the assumption that $\q'$ is massless.
\begin{subentry}
\iteme{= 0 :} no.
\iteme{= 1 :} yes, but it is possible, as before, to switch off individual
channels by the setting of \ttt{MDME} switches. Mass effects
are taken into account, in a crude fashion, by rejecting events
where kinematics becomes inconsistent when the $\q'$ mass is included.
\end{subentry}
\iteme{MSTP(11) :} (D = 1) use of electron parton distribution in
$\ee$ and $\ep$ interactions.
\begin{subentry}
\iteme{= 0 :} no, i.e.\ electron carries the whole beam energy.
\iteme{= 1 :} yes, i.e.\ electron carries only a fraction of beam
energy in agreement with next-to-leading electron parton-distribution
function, thereby including the effects of initial-state
bremsstrahlung.
\end{subentry}
\iteme{MSTP(12) :} (D = 0) use of $\e^-$ (`sea', i.e.\ from
$\e \to \gamma \to \e$), $\e^+$, quark and gluon distribution
functions inside an electron.
\begin{subentry}
\iteme{= 0 :} off.
\iteme{= 1 :} on, provided that \ttt{MSTP(11)} $\geq 1$.
Quark and gluon distributions
are obtained by numerical convolution of the photon content
inside an electron (as given by the bremsstrahlung spectrum of
\ttt{MSTP(11) = 1}) with the quark and gluon content inside a
photon. The required numerical precision is set by \ttt{PARP(14)}.
Since the need for numerical integration makes this option
somewhat more time-consuming than ordinary parton-distribution
evaluation, one should only use it when studying processes
where it is needed.
\itemc{Note:} for all traditional photoproduction/DIS physics this
option is superseded by the {\galep} option for
\ttt{PYINIT} calls, but can still be of use for some less standard
processes.
\end{subentry}
\iteme{MSTP(13) :} (D = 1) choice of $Q^2$ range over which electrons
are assumed to radiate photons; affects normalization of $\e^-$
(sea), $\e^+$, $\gamma$, quark and gluon distributions inside an
electron for \ttt{MSTP(12) = 1}.
\begin{subentry}
\iteme{= 1 :} range set by $Q^2$ argument of
parton-distribution-function call, i.e.\ by $Q^2$ scale of the hard
interaction. Therefore
parton distributions are proportional to $\ln(Q^2/m_e^2)$.
\iteme{= 2 :} range set by the user-determined $Q_{\mmax}^2$, given in
\ttt{PARP(13)}. Parton distributions are assumed to be proportional to
$\ln((Q_{\mmax}^2/m_e^2)(1-x)/x^2)$. This is normally most appropriate
for photoproduction, where the electron is supposed to go
undetected, i.e.\ scatter less than $Q_{\mmax}^2$.
\itemc{Note:} the choice of effective range is especially touchy for
the quark and gluon distributions. An (almost) on-the-mass-shell
photon has a VMD piece that dies away for a virtual photon. A simple
convolution of distribution functions does not take this into account
properly. Therefore the contribution from $Q$ values above the
$\rho$ mass should be suppressed. A choice of
$Q_{\mmax} \approx 1$~GeV is then appropriate for a
photoproduction limit description of physics. See also note for
\ttt{MSTP(12)}.
\end{subentry}
\iteme{MSTP(14) :} (D = 30) structure of incoming photon beam or
target. Historically, numbers up to 10 were set up for real photons,
and subsequent ones have been added also to allow for virtual photon
beams. The reason is that the earlier options specify e.g.\
direct$\times$VMD, summing over the possibilities of which photon is
direct and which VMD. This is allowed when the situation is symmetric,
i.e.\ for two incoming real photons, but not if one is virtual. Some of
the new options agree with previous ones, but are included to allow a
more consistent pattern. Further options above 25 have been added also
to include DIS processes.
\begin{subentry}
\iteme{= 0 :} a photon is assumed to be point-like (a direct photon),
i.e.\ can only interact in processes which explicitly contain the
incoming photon, such as $\f_i \gamma \to \f_i \g$ for $\gamma\p$
interactions. In $\gamma\gamma$ interactions both photons are
direct, i.e the main process is $\gamma \gamma \to \f_i \fbar_i$.
\iteme{= 1 :} a photon is assumed to be resolved, i.e.\ can only interact
through its constituent quarks and gluons, giving either high-$\pT$
parton--parton scatterings or low-$\pT$ events. Hard processes are
calculated with the use of the full photon parton distributions.
In $\gamma\gamma$ interactions both photons are resolved.
\iteme{= 2 :} a photon is assumed resolved, but only the VMD piece is
included in the parton distributions, which therefore mainly
are scaled-down versions of the $\rho^0 / \pi^0$ ones. Both high-$\pT$
parton--parton scatterings and low-$\pT$ events are allowed. In
$\gamma\gamma$ interactions both photons are VMD-like.
\iteme{= 3 :} a photon is assumed resolved, but only the anomalous
piece of the photon parton distributions is included. (This event
class is called either anomalous or GVMD; we will use both
interchangeably, though the former is more relevant for high-$\pT$
phenomena and the latter for low-$\pT$ ones.)
In $\gamma\gamma$
interactions both photons are anomalous.
\iteme{= 4 :} in $\gamma\gamma$ interactions one photon is direct
and the other resolved. A typical process is thus
$\f_i \gamma \to \f_i \g$. Hard processes are calculated with the use
of the full photon parton distributions for the resolved photon.
Both possibilities of which photon is direct are included, in event
topologies and in cross sections. This option cannot be used in
configurations with only one incoming photon.
\iteme{= 5 :} in $\gamma\gamma$ interactions one photon is direct
and the other VMD-like. Both possibilities of which photon is direct
are included, in event topologies and in cross sections. This option
cannot be used in configurations with only one incoming photon.
\iteme{= 6 :} in $\gamma\gamma$ interactions one photon is direct
and the other anomalous. Both possibilities of which photon is direct
are included, in event topologies and in cross sections. This option
cannot be used in configurations with only one incoming photon.
\iteme{= 7 :} in $\gamma\gamma$ interactions one photon is VMD-like
and the other anomalous. Only high-$\pT$ parton--parton scatterings
are allowed. Both possibilities of which photon is VMD-like are
included, in event topologies and in cross sections. This option
cannot be used in configurations with only one incoming photon.
\iteme{= 10 :} the VMD, direct and anomalous/GVMD components of the
photon are automatically mixed. For $\gamma\p$ interactions, this
means an automatic mixture of the three classes 0, 2 and 3 above
\cite{Sch93,Sch93a}, for $\gamma\gamma$ ones a mixture of the
six classes 0, 2, 3, 5, 6 and 7 above \cite{Sch94a}. Various
restrictions exist for this option, as discussed in section
\ref{sss:photoprodclass}.
\iteme{= 11 :} direct$\times$direct (see note 5);
intended for virtual photons.
\iteme{= 12 :} direct$\times$VMD (i.e.\ first photon direct, second VMD);
intended for virtual photons.
\iteme{= 13 :} direct$\times$anomalous; intended for virtual photons.
\iteme{= 14 :} VMD$\times$direct; intended for virtual photons.
\iteme{= 15 :} VMD$\times$VMD; intended for virtual photons.
\iteme{= 16 :} VMD$\times$anomalous; intended for virtual photons.
\iteme{= 17 :} anomalous$\times$direct; intended for virtual photons.
\iteme{= 18 :} anomalous$\times$VDM; intended for virtual photons.
\iteme{= 19 :} anomalous$\times$anomalous; intended for virtual photons.
\iteme{= 20 :} a mixture of the nine above components, 11--19, in the same
spirit as \ttt{= 10} provides a mixture for real gammas (or
a virtual gamma on a hadron). For gamma-hadron, this
option coincides with \ttt{= 10}.
\iteme{= 21 :} direct$\times$direct (see note 5).
\iteme{= 22 :} direct$\times$resolved.
\iteme{= 23 :} resolved$\times$direct.
\iteme{= 24 :} resolved$\times$resolved.
\iteme{= 25 :} a mixture of the four above components, offering a
simpler alternative to \ttt{= 20} in cases where the parton
distributions of the photon have not been split into VMD
and anomalous components. For $\gamma$-hadron, only two
components need be mixed.
\iteme{= 26 :} DIS$\times$VMD/$\p$.
\iteme{= 27 :} DIS$\times$anomalous.
\iteme{= 28 :} VMD/$\p$$\times$DIS.
\iteme{= 29 :} anomalous$\times$DIS.
\iteme{= 30 :} a mixture of all the 4 (for $\gamma^*\p$) or 13 (for
$\gamma^*\gamma^*$) components that are available, i.e.\ (the relevant
ones of) 11--19 and 26--29 above; is as \ttt{= 20} with the DIS
processes mixed in.
\itemc{Note 1:} the \ttt{MSTP(14)} options apply for a photon defined
by a \ttt{'gamma'} or {\galep} beam in the \ttt{PYINIT} call,
but not to those photons implicitly obtained in a \ttt{'lepton'} beam
with the \ttt{MSTP(12) = 1} option. This latter approach to resolved
photons is more primitive and is no longer recommended for QCD processes.
\itemc{Note 2:} for real photons our best understanding of how to mix
event classes is provided by the option 10 above, which also can be
obtained by combining three (for $\gamma\p$) or six (for $\gamma\gamma$)
separate runs. In a simpler alternative the VMD and anomalous classes are
joined into a single resolved class. Then $\gamma\p$ physics only requires
two separate runs, with 0 and 1, and $\gamma\gamma$ physics requires
three, with 0, 1 and 4.
\itemc{Note 3:} most of the new options from 11 onwards are not needed and
therefore not defined for $\e\p$ collisions. The recommended 'best' value
thus is \ttt{MSTP(14) = 30}, which also is the new default value.
\itemc{Note 4:} as a consequence of the appearance of new event classes,
the \ttt{MINT(122)} and \ttt{MSTI(9)} codes are not the same for
$\gamma^*\gamma^*$ events as for $\gamma\p$, $\gamma^*\p$ or
$\gamma\gamma$ ones. Instead the code is $3(i_1 - 1) + i_2$, where
$i$ is 1 for direct, 2 for VMD and 3 for anomalous/GVMD and indices
refer to the two incoming photons. For $\gamma^*\p$ code 4 is DIS,
and for $\gamma^* \gamma^*$ codes 10--13 corresponds to the \ttt{MSTP(14)}
codes 26--29. As before, \ttt{MINT(122)} and \ttt{MSTI(9)} are only
defined when several processes are to be mixed, not when generating one
at a time. Also the \ttt{MINT(123)} code is modified (not shown here).
\itemc{ Note 5:} the direct$\times$direct event class excludes lepton
pair production when run with the default \ttt{MSEL = 1} option (or
\ttt{MSEL = 2)}, in order not to confuse users. You can obtain lepton
pairs as well, e.g.\ by running with \ttt{MSEL = 0} and switching on the
desired processes by hand.
\itemc{ Note 6:} for all non-QCD processes, a photon is assumed unresolved
when \ttt{MSTP(14) =} 10, 20 or 25. In principle, both the resolved and
direct possibilities ought to be explored, but this mixing is not currently
implemented, so picking direct at least will explore one of the two
main alternatives rather than none. Resolved processes can be accessed
by the more primitive machinery of having a lepton beam and
\ttt{MSTP(12) = 1}.
\end{subentry}
\iteme{MSTP(15) :} (D = 0) possibility to modify the nature of the
anomalous photon component (as used with the appropriate \ttt{MSTP(14)}
options), in particular with respect to the scale choices and
cut-offs of hard processes. These options are mainly intended for
comparative studies and should not normally be touched. Some of the
issues are discussed in \cite{Sch93a}, while others have only been used
for internal studies and are undocumented.
\begin{subentry}
\iteme{= 0 :} none, i.e.\ the same treatment as for the VMD component.
\iteme{= 1 :} evaluate the anomalous parton distributions at a scale
$Q^2/$\ttt{PARP(17)}$^2$.
\iteme{= 2 :} as \ttt{= 1}, but instead of \ttt{PARP(17)} use either
\ttt{PARP(81)/PARP(15)} or \ttt{PARP(82)/PARP(15)}, depending on
\ttt{MSTP(82)} value.
\iteme{= 3 :} evaluate the anomalous parton distribution functions
of the photon as $f^{\gamma,\mrm{anom}}(x, Q^2, p_0^2) -
f^{\gamma,\mrm{anom}}(x, Q^2, r^2 Q^2)$ with $r = $\ttt{PARP(17)}.
\iteme{= 4 :} as \ttt{= 3}, but instead of \ttt{PARP(17)} use either
\ttt{PARP(81)/PARP(15)} or \ttt{PARP(82)/PARP(15)}, depending on
\ttt{MSTP(82)} value.
\iteme{= 5 :} use larger $\pTmin$ for the anomalous component than
for the VMD one, but otherwise no difference.
\end{subentry}
\iteme{MSTP(16) :} (D = 1) choice of definition of the fractional momentum
taken by a photon radiated off a lepton. Enters in the flux
factor for the photon rate, and thereby in cross sections.
\begin{subentry}
\iteme{= 0 :} $x$, i.e.\ energy fraction in the rest frame of the event.
\iteme{= 1 :} $y$, i.e.\ light-cone fraction.
\end{subentry}
\iteme{MSTP(17) :} (D = 4) possibility of a extra factor for processes
involving resolved virtual photons, to approximately take into account
the effects of longitudinal photons. Given on the form\\
$R = 1 + \mbox{\ttt{PARP(165)}} \, r(Q^2,\mu^2) \,
f_L(y,Q^2)/f_T(y,Q^2)$.\\
Here the 1 represents the basic transverse contribution,
\ttt{PARP(165)} is some arbitrary overall factor, and $f_L/f_T$
the (known) ratio of longitudinal to transverse photon
flux factors. The arbitrary function $r$ depends on the photon
virtuality $Q^2$ and the hard scale $\mu^2$ of the process.
See \cite{Fri00} for a discussion of the options.
\begin{subentry}
\iteme{= 0 :} No contribution, i.e.\ $r=0$.
\iteme{= 1 :} $r = 4 \mu^2 Q^2 / (\mu^2 + Q^2)^2$.
\iteme{= 2 :} $r = 4 Q^2 / (\mu^2 + Q^2)$.
\iteme{= 3 :} $r = 4 Q^2 / (m_{\rho}^2 + Q^2)$.
\iteme{= 4 :} $r = 4 m_V^2 Q^2 / (m_V^2 + Q^2)^2$, where $m_V$ is
the vector meson mass for VMD and $2\kT$ for GVMD states. Since there
is no $\mu$ dependence here (as well as for \ttt{= 3} and \ttt{= 5})
also minimum-bias cross sections are affected, where $\mu$ would be
vanishing. Currently the actual vector meson mass in the VMD case is
replaced by $m_{\rho}$, for simplicity.
\iteme{= 5 :} $r = 4 Q^2 / (m_V^2 + Q^2)$, with $m_V$ and comments
as above.
\iteme{Note:} for a photon given by the {\galep} option in
the \ttt{PYINIT} call, the $y$ spectrum is dynamically generated and
$y$ is thus known from event to event. For a photon beam in the
\ttt{PYINIT} call, $y$ is unknown from the onset, and has to be provided
by you if any longitudinal factor is to be included. So long
as these values, in \ttt{PARP(167)} and \ttt{PARP(168)}, are at their
default values, 0, it is assumed they have not been set and thus the
\ttt{MSTP(17)} and \ttt{PARP(165)} values are inactive.
\end{subentry}
\iteme{MSTP(18) :} (D = 3) choice of $\pTmin$ for direct processes.
\begin{subentry}
\iteme{= 1 :} same as for VMD and GVMD states, i.e.\ the $\pTmin(W^2)$
scale. Primarily intended for real photons.
\iteme{= 2 :} $\pTmin$ is chosen to be \ttt{PARP(15)}, i.e.\ the original
old behaviour proposed in \cite{Sch93,Sch93a}. In that case, also parton
distributions, jet cross sections and $\alphas$ values were dampened for
small $\pT$, so it may not be easy to obtain full backwards compatibility
with this option.
\iteme{= 3 :} as \ttt{= 1}, but if the $Q$ scale of the virtual photon is
above the VMD/GVMD $\pTmin(W^2)$, $\pTmin$ is chosen equal to $Q$.
This is part of the strategy to mix in DIS processes at $\pT$ below $Q$,
e.g.\ in \ttt{MSTP(14) = 30}.
\end{subentry}
\iteme{MSTP(19) :} (D = 4) choice of partonic cross section in the DIS
process 99.
\begin{subentry}
\iteme{= 0 :} QPM answer $4 \pi^2 \alphaem/Q^2 \,
\sum_{\q} e_{\q}^2 (x q(x,Q^2) + x \br{q}(x,Q^2))$
(with parton distributions frozen below the lowest $Q$
allowed in the parameterization). Note that this answer
is divergent for $Q^2 \to 0$ and thus violates gauge invariance.
\iteme{= 1 :} QPM answer is modified by a factor $Q^2/(Q^2 + m_{\rho}^2)$
to provide a finite cross section in the $Q^2 \to 0$ limit.
A minimal regularization recipe.
\iteme{= 2 :} QPM answer is modified by a factor
$Q^4/(Q^2 + m_{\rho}^2)^2$ to provide a vanishing cross section in
the $Q^2 \to 0$ limit. Appropriate if one assumes that the normal
photoproduction description gives the total cross section for $Q^2 = 0$,
without any DIS contribution.
\iteme{= 3 :} as \ttt{= 2}, but additionally suppression by a
parameterized factor $f(W^2,Q^2)$ (different for $\gamma^*\p$ and
$\gamma^*\gamma^*$) that avoids double-counting the direct-process
region where $\pT > Q$. Shower evolution for DIS events is then also
restricted to be at scales below $Q$, whereas evolution all
the way up to $W$ is allowed in the other options above.
\iteme{= 4 :} as \ttt{= 3}, but additionally include factor
$1/(1-x)$ for conversion from $F_2$ to $\sigma$. This is formally
required, but is only relevant for small $W^2$ and therefore often
neglected.
\end{subentry}
\iteme{MSTP(20) :} (D = 3) suppression of resolved (VMD or GVMD) cross
sections, introduced to compensate for an overlap with DIS processes in
the region of intermediate $Q^2$ and rather small $W^2$.
\begin{subentry}
\iteme{= 0 :} no; used as is.
\iteme{> 1 :} yes, by a factor
$(W^2/(W^2 + Q_1^2 + Q_2^2))^{\mbox{\ttt{MSTP(20)}}}$
(where $Q_i^2 = 0$ for an incoming hadron).
\iteme{Note:} the suppression factor is joined with the dipole
suppression stored in \ttt{VINT(317)} and \ttt{VINT(318)}.
\end{subentry}
\iteme{MSTP(21) :} (D = 1) nature of fermion--fermion scatterings
simulated in process 10 by $t$-channel exchange.
\begin{subentry}
\iteme{= 0 :} all off (!).
\iteme{= 1 :} full mixture of $\gammaZ$ neutral current and
$\W^{\pm}$ charged current.
\iteme{= 2 :} $\gamma$ neutral current only.
\iteme{= 3 :} $\Z^0$ neutral current only.
\iteme{= 4 :} $\gammaZ$ neutral current only.
\iteme{= 5 :} $\W^{\pm}$ charged current only.
\end{subentry}
\iteme{MSTP(22) :} (D = 0) special override of normal $Q^2$ definition
used for maximum of parton-shower evolution, intended for Deeply
Inelastic Scattering in lepton--hadron events, see section
\ref{ss:showrout}.
\iteme{MSTP(23) :} (D = 1) for Deeply Inelastic Scattering processes
(10 and 83), this option allows the $x$ and $Q^2$ of the original
hard scattering to be retained by the final electron when showers
are considered (with warnings as below; partly obsolete).
\begin{subentry}
\iteme{= 0 :} no correction procedure, i.e.\ $x$ and $Q^2$ of the
scattered electron differ from the originally generated $x$ and
$Q^2$.
\iteme{= 1 :} post facto correction, i.e.\ the change of electron
momentum, by initial and final QCD radiation, primordial $k_{\perp}$
and beam-remnant treatment, is corrected for by a shuffling of
momentum between the electron and hadron side in the final state.
Only process 10 is corrected, while process 83 is not.
\iteme{= 2 :} as \ttt{= 1}, except that both process 10 and 83 are
treated. This option is dangerous, especially for top, since it
may well be impossible to `correct' in process 83: the standard DIS
kinematics definitions are based on the assumption of massless quarks.
Therefore infinite loops are not excluded.
\itemc{Note 1:} the correction procedure will fail for a fraction of
the events, which are thus rejected (and new ones generated in their
place). The correction option is not unambiguous, and should
not be taken too seriously. For very small $Q^2$ values, the $x$
is not exactly preserved even after this procedure.
\itemc{Note 2:} this switch does not affect the recommended DIS
description obtained with a {\galep} beam/target in
\ttt{PYINIT}, where $x$ and $Q^2$ are always conserved.
\end{subentry}
\iteme{MSTP(25) :} (D = 0) angular decay correlations in Higgs decays
to $\W^+ \W^-$ or $\Z^0 \Z^0$ to four fermions \cite{Skj93}.
\begin{subentry}
\iteme{= 0 :} assuming the Higgs decay is pure scalar for
$\hrm^0$ and $\H^0$ and pure pseudoscalar for $\A^0$.
\iteme{= 1 :} assuming the Higgs decay is always pure scalar (CP-even).
\iteme{= 2 :} assuming the Higgs decay is always pure pseudoscalar
(CP-odd).
\iteme{= 3 :} assuming the Higgs decay is a mixture of the two
(CP-even and CP-odd), including the CP-violating interference term.
The parameter $\eta$, \ttt{PARP(25)} sets the strength of the
CP-odd admixture, with the interference term being proportional
to $\eta$ and the CP-odd one to $\eta^2$.
\itemc{Note :} since the decay of an $\A^0$ to $\W^+ \W^-$ or
$\Z^0 \Z^0$ is vanishing at the Born level, and no loop diagrams
are included, currently this switch is only relevant for $\hrm^0$ and
$\H^0$. It is mainly intended to allow `straw man' studies of
the quantum numbers of a Higgs state, decoupled from the issue of
branching ratios.
\end{subentry}
\iteme{MSTP(31) :} (D = 1) parameterization of total, elastic and
diffractive cross sections.
\begin{subentry}
\iteme{= 0 :} everything is to be set by you yourself in
the \ttt{PYINT7} common block. For photoproduction, additionally
you need to set \ttt{VINT(281)}. Normally you would set these
values once and for all before the \ttt{PYINIT} call, but if
you run with variable energies (see \ttt{MSTP(171)}) you can
also set it before each new \ttt{PYEVNT} call.
\iteme{= 1 :} Donnachie--Landshoff for total cross section
\cite{Don92}, and Schuler--Sj\"ostrand for elastic and diffractive
cross sections \cite{Sch94,Sch93a}.
\end{subentry}
\iteme{MSTP(32) :} (D = 8) $Q^2$ definition in hard scattering for
$2 \to 2$ processes. For resonance production $Q^2$ is always chosen
to be $\hat{s} = m_R^2$, where $m_R$ is the mass of the resonance.
For gauge boson scattering processes $VV \to VV$ the $\W$ or $\Z^0$
squared mass is used as scale in parton distributions. See
\ttt{PARP(34)} for a possibility to modify the choice below by
a multiplicative factor.\\
The newer options 6--10 are specifically intended for processes with
incoming virtual photons. These are ordered from a `minimal'
dependence on the virtualities to a `maximal' one, based on
reasonable kinematics considerations. The old default value
\ttt{MSTP(32) = 2} forms the starting point, with no dependence at
all, and the new default is some intermediate choice.
Notation is that $P_1^2$ and $P_2^2$ are the virtualities of the
two incoming particles, $\pT$ the transverse momentum of the
scattering process, and $m_3$ and $m_4$ the masses of the two
outgoing partons. For a direct photon, $P^2$ is the photon
virtuality and $x=1$. For a resolved photon, $P^2$ still refers
to the photon, rather than the unknown virtuality of the
reacting parton in the photon, and $x$ is the momentum fraction
taken by this parton.
\begin{subentry}
\iteme{= 1 :} $Q^2 = 2 \hat{s} \hat{t} \hat{u} / (\hat{s}^2 +
\hat{t}^2 + \hat{u}^2)$.
\iteme{= 2 :} $Q^2 = (m_{\perp 3}^2 + m_{\perp 4}^2)/2 =
\pT^2 + (m_3^2 + m_4^2)/2$.
\iteme{= 3 :} $Q^2 = \min(-\hat{t}, -\hat{u})$.
\iteme{= 4 :} $Q^2 = \hat{s}$.
\iteme{= 5 :} $Q^2 = -\hat{t}$.
\iteme{= 6 :} $Q^2 = (1 + x_1 P_1^2/\hat{s} + x_2 P_2^2/\hat{s})
(\pT^2 + m_3^2/2 + m_4^2/2)$.
\iteme{= 7 :} $Q^2 = (1 + P_1^2/\hat{s} + P_2^2/\hat{s})
(\pT^2 + m_3^2/2 + m_4^2/2)$.
\iteme{= 8 :} $Q^2 = \pT^2 + (P_1^2 + P_2^2 + m_3^2 + m_4^2)/2$.
\iteme{= 9 :} $Q^2 = \pT^2 + P_1^2 + P_2^2 +m_3^2 + m_4^2$.
\iteme{= 10 :} $Q^2 = s$ (the full energy-squared of the process).
\iteme{= 11 :} $Q^2 = (m_3 + m_4)^2/4$.
\iteme{= 12 :} $Q^2$ is set by the user as fixed numbers,
factorization scale in \ttt{PARP(193)} and renormalization scale
in \ttt{PARP(194)}.
\iteme{= 13 :} $Q^2 = \pT^2$, i.e.\ without any dependence on masses.
\itemc{Note:} options 6 and 7 are motivated by assuming that one
wants a scale that interpolates between $\hat{t}$ for small
$\hat{t}$ and $\hat{u}$ for small $\hat{u}$, such as
$Q^2 = - \hat{t}\hat{u}/(\hat{t}+\hat{u})$. When kinematics for
the $2 \to 2$ process is constructed as if an incoming
photon is massless when it is not, it gives rise to a
mismatch factor $1 + P^2/\hat{s}$ (neglecting the other
masses) in this $Q^2$ definition, which is then what is
used in option 7 (with the neglect of some small
cross-terms when both photons are virtual). When a
virtual photon is resolved, the virtuality of the
incoming parton can be anything from $xP^2$ and upwards.
So option 6 uses the smallest kinematically possible
value, while 7 is more representative of the typical
scale. Option 8 and 9 are more handwaving extensions of
the default option, with 9 specially constructed to
ensure that the $Q^2$ scale is always bigger than $P^2$.
\end{subentry}
\iteme{MSTP(33) :} (D = 0) inclusion of $K$ factors in hard
cross sections for parton--parton interactions (i.e.\ for
incoming quarks and gluons).
\begin{subentry}
\iteme{= 0 :} none, i.e.\ $K = 1$.
\iteme{= 1 :} a common $K$ factor is used, as stored in
\ttt{PARP(31)}.
\iteme{= 2 :} separate factors are used for ordinary
(\ttt{PARP(31)}) and colour annihilation graphs (\ttt{PARP(32)}).
\iteme{= 3 :} A $K$ factor is introduced by a shift in the
$\alphas$ $Q^2$ argument,
$\alphas = \alphas($\ttt{PARP(33)}$Q^2)$.
\end{subentry}
\iteme{MSTP(34) :} (D = 1) use of interference term in matrix
elements for QCD processes, see section \ref{sss:QCDjetclass}.
\begin{subentry}
\iteme{= 0 :} excluded (i.e.\ string-inspired matrix elements).
\iteme{= 1 :} included (i.e.\ conventional QCD matrix elements).
\itemc{Note:} for the option \ttt{MSTP(34) = 1}, i.e.\ interference
terms included, these terms are divided between the different
possible colour configurations according to the pole structure
of the (string-inspired) matrix elements for the different
colour configurations.
\end{subentry}
\iteme{MSTP(35) :} (D = 0) threshold behaviour for heavy-flavour
production, i.e.\ {\ISUB} = 81, 82, 84, 85, and also for $\Z$ and $\Z'$
decays. The non-standard options are mainly intended for top, but
can be used, with less theoretical reliability, also for charm and
bottom (for $\Z$ and $\Z'$ only top and heavier flavours
are affected). The threshold factors are given in
eqs.~(\ref{pp:threshenh}) and (\ref{pp:threshsup}).
\begin{subentry}
\iteme{= 0 :} na\"{\i}ve lowest-order matrix-element behaviour.
\iteme{= 1 :} enhancement or suppression close to threshold,
according to the colour structure of the process. The
$\alphas$ value appearing in the threshold factor (which is not
the same as the $\alphas$ of the lowest-order $2 \to 2$ process)
is taken to be fixed at the value given in \ttt{PARP(35)}. The
threshold factor used in an event is stored in \ttt{PARI(81)}.
\iteme{= 2 :} as \ttt{= 1}, but the $\alphas$ value appearing in
the threshold factor is taken to be running, with argument
$Q^2 = m_{\Q} \sqrt{ (\hat{m} - 2m_{\Q})^2 + \Gamma_{\Q}^2}$.
Here $m_{\Q}$ is the nominal heavy-quark mass, $\Gamma_{\Q}$ is
the width of the heavy-quark-mass distribution, and $\hat{m}$ is
the invariant mass of the heavy-quark pair. The $\Gamma_{\Q}$ value
has to be stored by you in \ttt{PARP(36)}. The regularization
of $\alphas$ at low $Q^2$ is given by \ttt{MSTP(36)}.
\end{subentry}
\iteme{MSTP(36) :} (D = 2) regularization of $\alphas$ in the
limit $Q^2 \to 0$ for the threshold factor obtainable in the
\ttt{MSTP(35) = 2} option; see \ttt{MSTU(115)} for a list of
the possibilities.
\iteme{MSTP(37) :} (D = 1) inclusion of running quark masses in
Higgs production ($\q \qbar \to \hrm^0$) and decay
($\hrm^0 \to \q \qbar$) couplings, obtained by calls to the
\ttt{PYMRUN} function. Also included for charged Higgs
and technipion production and decay.
\begin{subentry}
\iteme{= 0 :} not included, i.e.\ fixed quark masses are used
according to the values in the \ttt{PMAS} array.
\iteme{= 1 :} included, with running starting from the
value given in the \ttt{PMAS} array, at a $Q_0$
scale of \ttt{PARP(37)} times the quark mass itself,
up to a $Q$ scale given by the Higgs mass.
This option only works when $\alphas$ is allowed to run (so one can
define a $\Lambda$ value). Therefore it is only applied if additionally
\ttt{MSTP(2)} $\geq 1$.
\end{subentry}
\iteme{MSTP(38) :} (D = 5) handling of quark loop masses in the box
graphs $\g \g \to \gamma \gamma$ and $\g \g \to \g \gamma$, and in
the Higgs production loop graphs $\q \qbar \to \g \hrm^0$,
$\q \g \to \q \hrm^0$ and $\g \g \to \g \hrm^0$, and their
equivalents with $\H^0$ or $\A^0$ instead of $\hrm^0$.
\begin{subentry}
\iteme{= 0 :} for $\g \g \to \gamma \gamma$ and $\g \g \to \g \gamma$
the program will for each flavour automatically choose the massless
approximation for light quarks and the full massive formulae for heavy
quarks, with a dividing line between light and heavy quarks that depends
on the actual $\hat{s}$ scale. For Higgs production, all quark loop
contributions are included with the proper masses. This option is then
correct only in the Standard Model Higgs scenario, and should not be
used e.g. in the MSSM.
\iteme{$\geq$1 :} for $\g \g \to \gamma \gamma$ and $\g \g \to \g \gamma$
the program will use the massless approximation throughout, assuming the
presence of \ttt{MSTP(38)} effectively massless quark species (however,
at most 8). Normally one would use \ttt{= 5} for the inclusion of all quarks
up to bottom, and \ttt{= 6} to include top as well. For Higgs production,
the approximate expressions derived in the $m_{\t} \to \infty$ limit are
used, rescaled to match the correct $\g \g \to \hrm^0/\H^0/\A^0$ cross
sections. This procedure should work, approximately, also for non-standard
Higgs particles.
\itemc{Warning:} for \ttt{= 0}, numerical instabilities may arise
in $\g \g \to \gamma \gamma$ and $\g \g \to \g \gamma$ for scattering at
small angles. You are therefore recommended in this case to set
\ttt{CKIN(27)} and \ttt{CKIN(28)} so as to exclude the range of scattering
angles that is not of interest anyway. Numerical problems may also occur
for Higgs production with \ttt{= 0}, and additionally the lengthy expressions
make the code error-prone.
\end{subentry}
\iteme{MSTP(39) :} (D = 2) choice of $Q^2$ scale for parton distributions
and initial-state parton showers in processes $\g\g \to \Q \Qbar \hrm$
or $\q \qbar \to \Q \Qbar \hrm$.
\begin{subentry}
\iteme{= 1 :} $m_{\Q}^2$.
\iteme{= 2 :} $\max(m_{\perp\Q}^2,m_{\perp\Qbar}^2 ) =
m_{\Q}^2 + \max(p_{\perp\Q}^2 , p_{\perp\Qbar}^2)$.
\iteme{= 3 :} $m_{\hrm}^2$, where $m_{\hrm}$ is the actual Higgs mass of
the event, fluctuating from one event to the next.
\iteme{= 4 :} $\hat{s} = (p_{\hrm} + p_{\Q} + p_{\Qbar})^2$.
\iteme{= 5 :} $m_{\hrm}^2$, where $m_{\hrm}$ is the nominal, fixed
Higgs mass.
\iteme{= 6 :} $(m_3 + m_5)^2/4$.
\iteme{= 7 :} $(m_{\perp 3}^2 + m_{\perp 4}^2)/2$.
\iteme{= 8 :} set by the user as fixed numbers, factorization scale in
\ttt{PARP(193)} and renormalization scale in \ttt{PARP(194)}.
\end{subentry}
\iteme{MSTP(40) :} (D = 0) option for Coulomb correction in process 25,
$\W^+\W^-$ pair production, see \cite{Kho96}. The value of the Coulomb
correction factor for each event is stored in \ttt{VINT(95)}.
\begin{subentry}
\iteme{= 0 :} `no Coulomb'. Is the often-used reference point.
\iteme{= 1 :} `unstable Coulomb', gives the correct first-order
expression valid in the non-relativistic limit. Is the reasonable
option to use as a `best bet' description of LEP 2 physics.
\iteme{= 2 :} `second-order Coulomb' gives the correct second-order
expression valid in the non-relativistic limit. In principle
this is even better than \ttt{= 1}, but the differences are negligible
and computer time does go up because of the need for a
numerical integration in the weight factor.
\iteme{= 3 :} `dampened Coulomb', where the unstable Coulomb
expression has been modified by a $(1-\beta)^2$ factor in front of the
arctan term. This is intended as an alternative to \ttt{= 1} within the
band of our uncertainty in the relativistic limit.
\iteme{= 4 :} `stable Coulomb', i.e.\ effects are calculated as if
the $\W$'s were stable. Is incorrect, and mainly intended for comparison
purposes.
\itemc{Note :} unfortunately the $\W$'s at LEP 2 are not in the
non-relativistic limit, so the separation of Coulomb effects from other
radiative corrections is not gauge invariant. The options above should
therefore be viewed as indicative only, not as the ultimate answer.
\end{subentry}
\iteme{MSTP(41) :} (D = 2) master switch for all resonance decays
($\Z^0$, $\W^{\pm}$, $\t$, $\hrm^0$, $\Z'^0$, $\W'^{\pm}$, $\H^0$,
$\A^0$, $\H^{\pm}$, $\L_{\Q}$, $\R^0$, $\d^*$, $\u^*$, \ldots).
\begin{subentry}
\iteme{= 0 :} all off.
\iteme{= 1 :} all on.
\iteme{= 2 :} on or off depending on their individual \ttt{MDCY} values.
\itemc{Note:} also for \ttt{MSTP(41) = 1} it is possible to switch
off the decays of specific resonances by using the \ttt{MDCY(KC,1)}
switches in {\Py}. However, since the \ttt{MDCY} values are
overwritten in the \ttt{PYINIT} call when \ttt{MSTP(41) = 1} (or 0),
individual resonances should then be switched off after the \ttt{PYINIT}
call.
\itemc{Warning:} for top, leptoquark and other colour-carrying resonances
it is dangerous to switch off decays if one later on intends to let them
decay (with \ttt{PYEXEC}); see section \ref{sss:LQclass}.
\end{subentry}
\iteme{MSTP(42) :} (D = 1) mass treatment in $2 \to 2$ processes, where
the final-state resonances have finite width (see \ttt{PARP(41)}).
(Does not apply for the production of a single $s$-channel resonance,
where the mass appears explicitly in the cross section of the
process, and thus is always selected with width.)
\begin{subentry}
\iteme{= 0 :} particles are put on the mass shell.
\iteme{= 1 :} mass generated according to a Breit--Wigner.
\end{subentry}
\iteme{MSTP(43) :} (D = 3) treatment of $\Z^0/\gamma^*$ interference
in matrix elements. So far implemented in subprocesses 1, 15, 19, 22,
30 and 35; in other processes what is called a $\Z^0$ is really a
$\Z^0$ only, without the $\gamma^*$ piece.
\begin{subentry}
\iteme{= 1 :} only $\gamma^*$ included.
\iteme{= 2 :} only $\Z^0$ included.
\iteme{= 3 :} complete $\Z^0/\gamma^*$ structure (with
interference) included.
\end{subentry}
\iteme{MSTP(44) :} (D = 7) treatment of $\Z'^0/\Z^0/\gamma^*$
interference in matrix elements.
\begin{subentry}
\iteme{= 1 :} only $\gamma^*$ included.
\iteme{= 2 :} only $\Z^0$ included.
\iteme{= 3 :} only $\Z'^0$ included.
\iteme{= 4 :} only $\Z^0/\gamma^*$ (with interference) included.
\iteme{= 5 :} only $\Z'^0/\gamma^*$ (with interference) included.
\iteme{= 6 :} only $\Z'^0/\Z^0$ (with interference) included.
\iteme{= 7 :} complete $\Z'^0/\Z^0/\gamma^*$ structure
(with interference) included.
\end{subentry}
\iteme{MSTP(45) :} (D = 3) treatment of $\W\W \to \W\W$ structure
({\ISUB} = 77).
\begin{subentry}
\iteme{= 1 :} only $\W^+\W^+ \to \W^+\W^+$ and
$\W^-\W^- \to \W^-\W^-$ included.
\iteme{= 2 :} only $\W^+\W^- \to \W^+\W^-$ included.
\iteme{= 3 :} all charge combinations $\W\W \to \W\W$ included.
\end{subentry}
\iteme{MSTP(46) :} (D = 1) treatment of $VV \to V'V'$ structures
({\ISUB} = 71--77), where $V$ represents a longitudinal gauge boson.
\begin{subentry}
\iteme{= 0 :} only $s$-channel Higgs exchange included (where
existing). With this option, subprocesses 71--72 and 76--77
will essentially be equivalent to subprocesses 5 and 8,
respectively, with the proper decay channels (i.e.\ only $\Z^0\Z^0$
or $\W^+\W^-$) set via \ttt{MDME}.
The description obtained for subprocesses 5 and 8 in this case
is more sophisticated, however; see section \ref{sss:heavySMHclass}.
\iteme{= 1 :} all graphs contributing to $VV \to V'V'$
processes are included.
\iteme{= 2 :} only graphs not involving Higgs exchange
(either in $s$, $t$ or $u$ channel) are included; this option
then gives the na\"{\i}ve behaviour one would expect if no Higgs
exists, including unphysical unitarity violations at high energies.
\iteme{= 3 :} the strongly interacting Higgs-like model of
Dobado, Herrero and Terron \cite{Dob91} with Pad\'e unitarization.
Note that to use this option it is necessary to set the Higgs mass
to a large number like 20 TeV (i.e.\ \ttt{PMAS(25,1) = 20000}). The
parameter $\nu$ is stored in \ttt{PARP(44)}, but should not have
to be changed.
\iteme{= 4 :} as \ttt{= 3}, but with K-matrix unitarization \cite{Dob91}.
\iteme{= 5 :} the strongly interacting QCD-like model of
Dobado, Herrero and Terron \cite{Dob91} with Pad\'e unitarization.
The parameter $\nu$ is stored in \ttt{PARP(44)}, but should not
have to be changed. The effective techni-$\rho$ mass in this model
is stored in \ttt{PARP(45)}; by default it is 2054 GeV, which is
the expected value for three technicolors, based on scaling up
the ordinary $\rho$ mass appropriately.
\iteme{= 6 :} as \ttt{= 5}, but with K-matrix unitarization \cite{Dob91}.
While \ttt{PARP(45)} still is a parameter of the model, this type
of unitarization does not give rise to a resonance at a mass of
\ttt{PARP(45)}.
\end{subentry}
\iteme{MSTP(47) :} (D = 1) (C) angular orientation of decay products
of resonances ($\Z^0$, $\W^{\pm}$, $\t$, $\hrm^0$, $\Z'^0$, $\W'^{\pm}$,
etc.), either when produced singly or in pairs (also
from an $\hrm^0$ decay), or in combination with a single quark,
gluon or photon.
\begin{subentry}
\iteme{= 0 :} independent decay of each resonance, isotropic in c.m.\
frame of the resonance.
\iteme{= 1 :} correlated decay angular distributions according to
proper matrix elements, to the extent these are implemented.
\end{subentry}
\iteme{MSTP(48) :} (D = 0) (C) switch for the treatment of
$\gamma^*/\Z^0$ decay for process 1 in $\ee$ events.
\begin{subentry}
\iteme{= 0 :} normal machinery.
\iteme{= 1 :} if the decay of the $\Z^0$ is to either of the five
lighter quarks, $\d$, $\u$, $\s$, $\c$ or $\b$, the special treatment
of $\Z^0$ decay is accessed, including matrix element options,
according to section \ref{ss:eematrix}.\\
This option is based on the machinery of the \ttt{PYEEVT} and associated
routines when it comes to the description of QCD multijet structure
and the angular orientation of jets, but relies on the normal
\ttt{PYEVNT} machinery for everything else: cross section calculation,
initial-state photon radiation, flavour composition of decays
(i.e.\ information on channels allowed), etc.\\
The initial state has to be $\ee$; forward-backward asymmetries would
not come out right for quark-annihilation production of the
$\gamma^*/\Z^0$ and therefore the machinery defaults to the standard
one in such cases.\\
You can set the behaviour for the \ttt{MSTP(48)} option using the normal
matrix element related switches. This especially means \ttt{MSTJ(101)} for
the selection of first- or second-order matrix elements (\ttt{= 1} and
\ttt{= 2}, respectively). Further selectivity is obtained with the switches
and parameters \ttt{MSTJ(102)} (for the angular orientation part only),
\ttt{MSTJ(103)} (except the production threshold factor part),
\ttt{MSTJ(106)}, \ttt{MSTJ(108) - MSTJ(111)}, \ttt{PARJ(121)},
\ttt{PARJ(122)}, and \ttt{PARJ(125) - PARJ(129)}.
Information can be read from \ttt{MSTJ(120)}, \ttt{MSTJ(121)},
\ttt{PARJ(150)}, \ttt{PARJ(152) - PARJ(156)}, \ttt{PARJ(168)},
\ttt{PARJ(169)}, \ttt{PARJ(171)}.\\
The other $\ee$ switches and parameters should not be touched. In most
cases they are simply not accessed, since the related part is handled
by the \ttt{PYEVNT} machinery instead. In other cases they could give
incorrect or misleading results. Beam polarization as set by
\ttt{PARJ(131) - PARJ(134)}, for instance, is only included for the
angular orientation, but is missing for the cross section information.
\ttt{PARJ(123)} and \ttt{PARJ(124)} for the $\Z^0$ mass and width are
set in the \ttt{PYINIT} call based on the input mass and calculated
widths.\\
The cross section calculation is unaffected by the matrix element
machinery. Thus also for negative \ttt{MSTJ(101)} values, where only
specific jet multiplicities are generated, the \ttt{PYSTAT} cross
section is the full one.
\end{subentry}
\iteme{MSTP(49) :} (D = 1) assumed variation of the Higgs width to massive
gauge boson pairs, i.e.\ $\W^+\W^-$, $\Z^0\Z^0$ and $\W^{\pm}\Z^0$, as a
function of the actual mass $\hat{m} = \sqrt{\hat{s}}$ and the nominal
mass $m_{\hrm}$. The switch applies both to $\hrm^0$, $\H^0$, $\A^0$ and
$\H^{\pm}$ decays.
\begin{subentry}
\iteme{= 0 :} the width is proportional to $\hat{m}^3$; thus the high-mass
tail of the Breit--Wigner is enhanced.
\iteme{= 1 :} the width is proportional to $m_{\hrm}^2 \hat{m}$. For
a fixed Higgs mass $m_{\hrm}$ this means a width variation across the
Breit--Wigner more in accord with other resonances (such as the $\Z^0$).
This alternative gives more emphasis to the low-mass tail, where the
parton distributions are peaked (for hadron colliders). This option is
favoured by resummation studies \cite{Sey95a}.
\itemc{Note 1:} the partial width of a Higgs to a fermion pair is always
taken to be proportional to the actual Higgs mass $\hat{m}$, irrespectively
of \ttt{MSTP(49)}. Also the width to a gluon or photon pair (via loops)
is unaffected.
\itemc{Note 2:} this switch does not affect processes 71--77, where a
fixed Higgs width is used in order to control cancellation of divergences.
\end{subentry}
\iteme{MSTP(50) :} (D = 0) Switch to allow or not longitudinally polarized
incoming beams, with the two polarizations stored in \ttt{PARJ(131)} and
\ttt{PARJ(132)}, respectively. Most cross section expressions with
polarization reduce to the unpolarized behaviour for the default
\ttt{PARJ(131) = PARJ(132) = 0.}, and then this switch is superfluous
and not implemented. Currently \ttt{MSTP(50)} is only used in process 25,
$\f \fbar \to \W^+ \W^-$, for reasons explained in section
\ref{ss:polarization}.
\begin{subentry}
\iteme{= 0 :} no polarization effects, no matter what \ttt{PARJ(131)}
and \ttt{PARJ(132)} values are set.
\iteme{= 1 :} include polarization information in the cross section of
the process and for angular correlations.
\end{subentry}
\iteme{MSTP(51) :} (D = 7) choice of proton parton-distribution set;
see also \ttt{MSTP(52)}.
\begin{subentry}
\iteme{= 1 :} CTEQ 3L (leading order).
\iteme{= 2 :} CTEQ 3M ($\br{\mrm{MS}}$).
\iteme{= 3 :} CTEQ 3D (DIS).
\iteme{= 4 :} GRV 94L (leading order).
\iteme{= 5 :} GRV 94M ($\br{\mrm{MS}}$).
\iteme{= 6 :} GRV 94D (DIS).
\iteme{= 7 :} CTEQ 5L (leading order).
\iteme{= 8 :} CTEQ 5M1 ($\br{\mrm{MS}}$; slightly updated version of
CTEQ 5M).
\iteme{= 11 :} GRV 92L (leading order).
\iteme{= 12 :} EHLQ set 1 (leading order; 1986 updated version).
\iteme{= 13 :} EHLQ set 2 (leading order; 1986 updated version).
\iteme{= 14 :} Duke--Owens set 1 (leading order).
\iteme{= 15 :} Duke--Owens set 2 (leading order).
\iteme{= 16 :} simple ansatz with all parton distributions of the
form $c/x$, with $c$ some constant; intended for internal debug use
only.
\itemc{Note 1:} distributions 11--15 are obsolete and should not be
used for current physics studies. They are only implemented to have
some sets in common between \Py~5 and 6, for cross-checks.
\itemc{Note 2:} since all parameterizations have some region of
applicability, the parton distributions are assumed frozen below
the lowest $Q^2$ covered by the parameterizations. In some cases,
evolution is also frozen above the maximum $Q^2$.
\end{subentry}
\iteme{MSTP(52) :} (D = 1) choice of proton
parton-distribution-function library.
\begin{subentry}
\iteme{= 1 :} the internal {\Py} one, with parton distributions
according to the \ttt{MSTP(51)} above.
\iteme{= 2 :} the \tsc{Pdflib} one \cite{Plo93}, with the
\tsc{Pdflib} (version 4) \ttt{NGROUP} and \ttt{NSET} numbers to be
given as \ttt{MSTP(51) = 1000}$\times$\ttt{NGROUP + NSET}, or
similarly for the \tsc{LHAPDF} one \cite{Gie02}.
\itemc{Note 1:} to make use of option 2, it is necessary to link
\tsc{Pdflib}/\tsc{LHAPDF}. Additionally, on most computers, the three
dummy routines \ttt{PDFSET}, \ttt{STRUCTM} and (for virtual photons)
\ttt{STRUCTP} at the end of the {\Py} file should be
removed or commented out.
\itemc{Warning:} for external parton distribution libraries,
{\Py} does not check whether \ttt{MSTP(51)} corresponds to a
valid code, or if special $x$ and $Q^2$ restrictions exist
for a given set, such that crazy values could be returned.
This puts an extra responsibility on you.
\itemc{Note 2:} when \tsc{Pdflib}/\tsc{LHAPDF} is used, {\Py} can
initialize either with a four- or a five-flavour $\Lambda$, depending
on how \ttt{NFL} in the \ttt{/W50511/} commonblock is set, extracting
either \ttt{QCDL4} or \ttt{QCDL5} from the \ttt{/W50512/} commonblock.
\end{subentry}
\iteme{MSTP(53) :} (D = 3) choice of pion parton-distribution set;
see also \ttt{MSTP(54)}.
\begin{subentry}
\iteme{= 1 :} Owens set 1.
\iteme{= 2 :} Owens set 2.
\iteme{= 3 :} GRV LO (updated version).
\end{subentry}
\iteme{MSTP(54) :} (D = 1) choice of pion parton-distribution-function
library.
\begin{subentry}
\iteme{= 1 :} the internal {\Py} one, with parton distributions
according to the \ttt{MSTP(53)} above.
\iteme{= 2 :} the \tsc{Pdflib} one \cite{Plo93}, with the
\tsc{Pdflib} (version 4) \ttt{NGROUP} and \ttt{NSET} numbers to be
given as \ttt{MSTP(53) = 1000}$\times$\ttt{NGROUP + NSET}, or
similarly for the \tsc{LHAPDF} one \cite{Gie02}.
\itemc{Note:} to make use of option 2, it is necessary to link
\tsc{Pdflib}/\tsc{LHAPDF}. Additionally, on most computers, the three
dummy routines \ttt{PDFSET}, \ttt{STRUCTM} and \ttt{STRUCTP} at the end
of the {\Py} file should be removed or commented out.
\itemc{Warning:} for external parton distribution libraries,
{\Py} does not check whether \ttt{MSTP(53)} corresponds to a valid
code, or if special $x$ and $Q^2$ restrictions exist for a given
set, such that crazy values could be returned. This puts an extra
responsibility on you.
\end{subentry}
\iteme{MSTP(55)} : (D = 5) choice of the parton-distribution
set of the photon; see also \ttt{MSTP(56)} and \ttt{MSTP(60)}.
\begin{subentry}
\iteme{= 1 :} Drees--Grassie.
\iteme{= 5 :} SaS 1D (in DIS scheme, with $Q_0=0.6$~GeV).
\iteme{= 6 :} SaS 1M (in {\MSbar} scheme, with $Q_0=0.6$~GeV).
\iteme{= 7 :} SaS 2D (in DIS scheme, with $Q_0=2$~GeV).
\iteme{= 8 :} SaS 2M (in {\MSbar} scheme, with $Q_0=2$~GeV).
\iteme{= 9 :} SaS 1D (in DIS scheme, with $Q_0=0.6$~GeV).
\iteme{= 10 :} SaS 1M (in {\MSbar} scheme, with $Q_0=0.6$~GeV).
\iteme{= 11 :} SaS 2D (in DIS scheme, with $Q_0=2$~GeV).
\iteme{= 12 :} SaS 2M (in {\MSbar} scheme, with $Q_0=2$~GeV).
\itemc{Note 1:} sets 5--8 use the parton distributions of the respective
set, and nothing else. These are appropriate for most applications, e.g.\
jet production in $\gamma\p$ and $\gamma\gamma$ collisions. Sets 9--12
instead are appropriate for $\gamma^*\gamma$ processes, i.e.\ DIS
scattering on a photon, as measured in $F_2^{\gamma}$. Here the anomalous
contribution for $\c$ and $\b$ quarks are handled by the Bethe-Heitler
formulae, and the direct term is artificially lumped with the anomalous
one, so that the event simulation more closely agrees with what will be
experimentally observed in these processes. The agreement with the
$F_2^{\gamma}$ parameterization is still not perfect, e.g.\ in the treatment
of heavy flavours close to threshold.
\itemc{Note 2:} sets 5--12 contain both VMD pieces and anomalous pieces,
separately parameterized. Therefore the respective piece is automatically
called, whatever \ttt{MSTP(14)} value is used to select only a part of the
allowed photon interactions. For other sets (set 1 above or
\tsc{Pdflib}/\tsc{LHAPDF} sets), usually there is no corresponding
subdivision. Then an option like \ttt{MSTP(14) = 2} (VMD part of photon
only) is based on a rescaling of the pion distributions, while
\ttt{MSTP(14) = 3} gives the SaS anomalous parameterization.
\itemc{Note 3:} formally speaking, the $k_0$ (or $p_0$) cut-off in
\ttt{PARP(15)} need not be set in any relation to the $Q_0$ cut-off
scales used by the various parameterizations. Indeed, due to the
familiar scale choice ambiguity problem, there could well be some offset
between the two. However, unless you know what you are doing, it is
recommended that you let the two agree, i.e.\ set
\ttt{PARP(15) = 0.6} for the SaS 1 sets and \ttt{= 2.} for the SaS 2 sets.
\end{subentry}
\iteme{MSTP(56) :} (D = 1) choice of photon parton-distribution-function
library.
\begin{subentry}
\iteme{= 1 :} the internal {\Py} one, with parton distributions
according to the \ttt{MSTP(55)} above.
\iteme{= 2 :} the \tsc{Pdflib} one \cite{Plo93}, with the
\tsc{Pdflib} (version 4) \ttt{NGROUP} and \ttt{NSET} numbers to be
given as \ttt{MSTP(55) = 1000}$\times$\ttt{NGROUP + NSET}, or
similarly for the \tsc{LHAPDF} one \cite{Gie02}.
When the VMD and anomalous parts of the photon are split,
like for \ttt{MSTP(14) = 10}, it is necessary to specify pion set to be
used for the VMD component, in \ttt{MSTP(53)} and \ttt{MSTP(54)},
while \ttt{MSTP(55)} here is irrelevant.
\iteme{= 3 :} when the parton distributions of the anomalous photon
are requested, the homogeneous solution is provided, evolved from a
starting value \ttt{PARP(15)} to the requested $Q$ scale. The homogeneous
solution is normalized so that the net momentum is unity,
i.e.\ any factors of $\alphaem/2\pi$ and charge have been left out.
The flavour of the original $\q$ is given in \ttt{MSTP(55)} (1, 2, 3, 4
or 5 for $\d$, $\u$, $\s$, $\c$ or $\b$); the value 0 gives a mixture
according to squared charge, with the exception that $\c$ and $\b$
are only allowed above the respective mass threshold ($Q > m_{\q}$).
The four-flavour $\Lambda$ value is assumed given in \ttt{PARP(1)};
it is automatically recalculated for 3 or 5 flavours at
thresholds. This option is not intended for standard event
generation, but is useful for some theoretical studies.
\itemc{Note:} to make use of option 2, it is necessary to link
\tsc{Pdflib}/\tsc{LHAPDF}. Additionally, on most computers, the three
dummy routines \ttt{PDFSET}, \ttt{STRUCTM} and \ttt{STRUCTP} at the
end of the {\Py} file should be removed or commented out.
\itemc{Warning 1:} for external parton-distribution libraries, {\Py}
does not check whether \ttt{MSTP(55)} corresponds to a valid code,
or if special $x$ and $Q^2$ restrictions exist for a given set,
such that crazy values could be returned. This puts an extra
responsibility on you.
\itemc{Warning 2:} so much of the machinery for virtual photons is
based on a subdivision of the photon according to the SaS prescription
that a usage of \tsc{Pdflib} cannot be recommended for such; in some
cases unphysical results may arise from mismatches between what
\tsc{Pdflib} delivers and what is assumed internally.
\end{subentry}
\iteme{MSTP(57) :} (D = 1) choice of $Q^2$ dependence in
parton-distribution functions.
\begin{subentry}
\iteme{= 0 :} parton distributions are evaluated at nominal lower
cut-off value $Q_0^2$, i.e.\ are made $Q^2$-independent.
\iteme{= 1 :} the parameterized $Q^2$ dependence is used.
\iteme{= 2 :} the parameterized parton-distribution behaviour is kept
at large $Q^2$ and $x$, but modified at small $Q^2$ and/or $x$,
so that parton distributions vanish in the limit $Q^2 \to 0$ and
have a theoretically motivated small-$x$ shape \cite{Sch93a}.
This option is only valid for the $\p$ and $\n$. It is obsolete within
the current {\galep} framework.
\iteme{= 3 :} as \ttt{= 2}, except that also the $\pi^{\pm}$ is modified
in a corresponding manner. A VMD photon is not mapped to a pion, but is
treated with the same photon parton distributions as for other
\ttt{MSTP(57)} values, but with properly modified behaviour for small
$x$ or $Q^2$. This option is obsolete within the current {\galep}
framework.
\end{subentry}
\iteme{MSTP(58) :} (D = min(5, 2$\times$\ttt{MSTP(1)})) maximum number of
quark flavours used in parton distributions, and thus also for
initial-state space-like showers. If some distributions (notably $\t$)
are absent in the parameterization selected in \ttt{MSTP(51)}, these
are obviously automatically excluded.
\iteme{MSTP(59) :} (D = 1) choice of electron-inside-electron parton
distribution.
\begin{subentry}
\iteme{= 1 :} the recommended standard for LEP 1, next-to-leading
exponentiated, see \cite{Kle89}, p. 34.
\iteme{= 2 :} the recommended `$\beta$' scheme for LEP 2, also
next-to-leading exponentiated, see \cite{Bee96}, p. 130.
\end{subentry}
\iteme{MSTP(60) :} (D = 7) extension of the SaS real-photon distributions to
off-shell photons, especially for the anomalous component. See \cite{Sch96}
for an explanation of the options. The starting point is the expression in
eq.~(\ref{eq:decompvirtga}), which requires a numerical integration of the
anomalous component, however, and therefore is not convenient. Approximately,
the dipole damping factor can be removed and compensated by a suitably
shifted lower integration limit, whereafter the integral simplifies.
Different `goodness' criteria for the choice of the shifted lower
limit is represented by the options 2--7 below.
\begin{subentry}
\iteme{= 1 :} dipole dampening by integration; very time-consuming.
\iteme{= 2 :} $P_0^2 = \max( Q_0^2, P^2 )$.
\iteme{= 3 :} ${P'}_0^2 = Q_0^2 + P^2$.
\iteme{= 4 :} $P_{\mrm{eff}}$ that preserves momentum sum.
\iteme{= 5 :} $P_{\mrm{int}}$ that preserves momentum and average
evolution range.
\iteme{= 6 :} $P_{\mrm{eff}}$, matched to $P_0$ in $P^2 \to Q^2$ limit.
\iteme{= 7 :} $P_{\mrm{int}}$, matched to $P_0$ in $P^2 \to Q^2$ limit.
\end{subentry}
\iteme{MSTP(61) :} (D = 2) (C) master switch for initial-state QCD and
QED radiation.
\begin{subentry}
\iteme{= 0 :} off.
\iteme{= 1 :} on.
\iteme{= 1 :} on for QCD radiation in hadronic events and QED
radiation in leptonic ones.
\iteme{= 2 :} on for QCD and QED radiation in hadronic events and
QED radiation in leptonic ones.
\end{subentry}
\iteme{MSTP(62) - MSTP(70) :} (C) further switches for initial-state
radiation, see section \ref{ss:showrout}.
\iteme{MSTP(71) :} (D = 1) (C) master switch for final-state QCD and
QED radiation.
\begin{subentry}
\iteme{= 0 :} off.
\iteme{= 1 :} on.
\itemc{Note:} additional switches (e.g.\ for conventional/coherent
showers) are available in \ttt{MSTJ(38) - MSTJ(50)} and
\ttt{PARJ(80) - PARJ(90)}, see section \ref{ss:showrout}.
\end{subentry}
\iteme{MSTP(72):} (C) further switch for initial-state
radiation, see section \ref{ss:showrout}.
\iteme{MSTP(81) :} (D = 1) master switch for multiple interactions.
\begin{subentry}
\iteme{= 0 :} off.
\iteme{= 1 :} on.
\end{subentry}
\iteme{MSTP(82) - MSTP(86) :} further switches for multiple
interactions, see section \ref{ss:multintpar}.
\iteme{MSTP(91) - MSTP(95) :} switches for beam-remnant treatment,
see section \ref{ss:multintpar}.
\iteme{MSTP(101) :} (D = 3) (C) structure of diffractive system.
\begin{subentry}
\iteme{= 1 :} forward moving diquark + interacting quark.
\iteme{= 2 :} forward moving diquark + quark joined via interacting
gluon (`hairpin' configuration).
\iteme{= 3 :} a mixture of the two options above, with a fraction
\ttt{PARP(101)} of the former type.
\end{subentry}
\iteme{MSTP(102) :} (D = 1) (C) decay of a $\rho^0$ meson produced by
`elastic' scattering of an incoming $\gamma$, as in
$\gamma \p \to \rho^0 \p$, or the same with the hadron diffractively
excited.
\begin{subentry}
\iteme{= 0 :} the $\rho^0$ is allowed to decay isotropically, like
any other $\rho^0$.
\iteme{= 1 :} the decay $\rho^0 \to \pi^+ \pi^-$ is done with an
angular distribution proportional to $\sin^2 \theta$ in its rest frame,
where the $z$ axis is given by the direction of motion of the
$\rho^0$. The $\rho^0$ decay is then done as part of the hard process,
i.e.\ also when \ttt{MSTP(111) = 0}.
\end{subentry}
\iteme{MSTP(110) :} (D = 0) switch to allow some or all resonance widths
to be modified by the factor \ttt{PARP(110)}. This is not intended for
serious physics studies. The main application is rather to generate
events with an artificially narrow resonance width in order to study
the detector-related smearing effects on the mass resolution.
\begin{subentry}
\iteme{> 0 :} rescale the particular resonance with \ttt{KF = MSTP(110)}.
If the resonance has an antiparticle, this one is affected as well.
\iteme{= -1 :} rescale all resonances, except $\t$, $\tbar$, $\Z^0$ and
$\W^{\pm}$.
\iteme{= -2 :} rescale all resonances.
\itemc{Warning:} only resonances with a width evaluated by \ttt{PYWIDT}
are affected, and preferentially then those with \ttt{MWID} value 1 or 3.
For other resonances the appearance of effects or not depends on how the
cross sections have been implemented. So it is important to check that
indeed the mass distribution is affected as expected. Also beware that,
if a sequential decay chain is involved, the scaling may become more
complicated. Furthermore, depending on implementational details, a cross
section may or may not scale with \ttt{PARP(110)} (quite apart from
differences related to the convolution with parton distributions etc.).
All in all, it is then an option to be used only with open eyes, and for
very specific applications.
\end{subentry}
\iteme{MSTP(111) :} (D = 1) (C) master switch for fragmentation
and decay, as obtained with a \ttt{PYEXEC} call.
\begin{subentry}
\iteme{= 0 :} off.
\iteme{= 1 :} on.
\iteme{= -1 :} only choose kinematical variables for hard scattering,
i.e.\ no jets are defined. This is useful, for instance, to calculate
cross sections (by Monte Carlo integration) without wanting
to simulate events; information obtained with \ttt{PYSTAT(1)}
will be correct.
\end{subentry}
\iteme{MSTP(112) :} (D = 1) (C) cuts on partonic events; only affects
an exceedingly tiny fraction of events. Normally this concerns what
happens in the \ttt{PYPREP} routine, if a colour singlet subsystem
has a very small invariant mass and attempts to collapse it to a single
particle fail, see section \ref{sss:smallmasssystem}.
\begin{subentry}
\iteme{= 0 :} no cuts (can be used only with independent
fragmentation, at least in principle).
\iteme{= 1 :} string cuts (as normally required for fragmentation).
\end{subentry}
\iteme{MSTP(113) :} (D = 1) (C) recalculation of energies of partons
from their momenta and masses, to be done immediately before
and after fragmentation, to partly compensate for some numerical
problems appearing at high energies.
\begin{subentry}
\iteme{= 0 :} not performed.
\iteme{= 1 :} performed.
\end{subentry}
\iteme{MSTP(115) :} (D = 0) (C) choice of colour rearrangement scenario
for process 25, $\W^+\W^-$ pair production, when both $\W$'s decay
hadronically. (Also works for process 22, $\Z^0\Z^0$ production,
except when the $\Z$'s are allowed to fluctuate to very small masses.)
See section \ref{sss:reconnect} for details.
\begin{subentry}
\iteme{= 0 :} no reconnection.
\iteme{= 1 :} scenario I, reconnection inspired by a type I superconductor,
with the reconnection probability related to the overlap volume in
space and time between the $\W^+$ and $\W^-$ strings. Related parameters
are found in \ttt{PARP(115) - PARP(119)}, with \ttt{PARP(117)} of special
interest.
\iteme{= 2 :} scenario II, reconnection inspired by a type II
superconductor, with reconnection possible when two string
cores cross. Related parameter in \ttt{PARP(115)}.
\iteme{= 3 :} scenario II', as model II but with the additional
requirement that a reconnection will only occur if the
total string length is reduced by it.
\iteme{= 5 :} the GH scenario, where the reconnection can occur that
reduces the total string length ($\lambda$ measure) most.
\ttt{PARP(120)} gives the fraction of such event where a
reconnection is actually made; since almost all events
could allow a reconnection that would reduce the string
length, \ttt{PARP(120)} is almost the same as the reconnection
probability.
\iteme{= 11 :} the intermediate scenario, where a reconnection is
made at the `origin' of events, based on the subdivision
of all radiation of a $\q\qbar$ system as coming either from
the $\q$ or the $\qbar$. \ttt{PARP(120)} gives the assumed probability
that a reconnection will occur. A somewhat simpleminded
model, but not quite unrealistic.
\iteme{= 12 :} the instantaneous scenario, where a reconnection is
allowed to occur before the parton showers, and showering
is performed inside the reconnected systems with maximum
virtuality set by the mass of the reconnected systems.
\ttt{PARP(120)} gives the assumed probability that a reconnection
will occur. Is completely unrealistic, but useful as an
extreme example with very large effects.
\end{subentry}
\iteme{MSTP(121) :} (D = 0) calculation of kinematics selection
coefficients and differential cross section maxima for
included (by you or default) subprocesses.
\begin{subentry}
\iteme{= 0 :} not known; to be calculated at initialization.
\iteme{= 1 :} not known; to be calculated at initialization;
however, the maximum value then obtained is to be multiplied by
\ttt{PARP(121)} (this may be useful if a violation factor has
been observed in a previous run of the same kind).
\iteme{= 2 :} known; kinematics selection coefficients stored
by you in \ttt{COEF(ISUB,J)} (\ttt{J} = 1--20) in common block
\ttt{PYINT2} and maximum of the corresponding differential
cross section times Jacobians in \ttt{XSEC(ISUB,1)} in
common block \ttt{PYINT5}. This is to be done for each included
subprocess {\ISUB} before initialization, with the sum of all
\ttt{XSEC(ISUB,1)} values, except for {\ISUB} = 95, stored in
\ttt{XSEC(0,1)}.
\end{subentry}
\iteme{MSTP(122) :} (D = 1) initialization and differential
cross section maximization print-out. Also, less importantly, level
of information on where in phase space a cross section maximum has
been violated during the run.
\begin{subentry}
\iteme{= 0 :} none.
\iteme{= 1 :} short message at initialization; only when an error
(i.e.\ not a warning) is generated during the run.
\iteme{= 2 :} detailed message, including full maximization., at
initialization; always during run.
\end{subentry}
\iteme{MSTP(123) :} (D = 2) reaction to violation of maximum
differential cross section or to occurence of negative differential
cross sections (except when allowed for external processes, i.e.\
when \ttt{IDWTUP < 0}).
\begin{subentry}
\iteme{= 0 :} stop generation, print message.
\iteme{= 1 :} continue generation, print message for each
subsequently larger violation.
\iteme{= 2 :} as \ttt{= 1}, but also increase value of maximum.
\end{subentry}
\iteme{MSTP(124) :} (D = 1) (C) frame for presentation of event.
\begin{subentry}
\iteme{= 1 :} as specified in \ttt{PYINIT}.
\iteme{= 2 :} c.m.\ frame of incoming particles.
\iteme{= 3 :} hadronic c.m.\ frame for DIS events, with warnings
as given for \ttt{PYFRAM}.
\end{subentry}
\iteme{MSTP(125) :} (D = 1) (C) documentation of partonic process,
see section \ref{sss:PYrecord} for details.
\begin{subentry}
\iteme{= 0 :} only list ultimate string/particle configuration.
\iteme{= 1 :} additionally list short summary of the hard process.
\iteme{= 2 :} list complete documentation of intermediate steps of
parton-shower evolution.
\end{subentry}
\iteme{MSTP(126) :} (D = 100) number of lines at the beginning of event
record that are reserved for event-history information; see section
\ref{sss:PYrecord}. This value should never be reduced, but may be
increased at a later date if more complicated processes are included.
\iteme{MSTP(127) :} (D = 0) possibility to continue run even if none
of the requested processes have non-vanishing cross sections.
\begin{subentry}
\iteme{= 0 :} no, the run will be stopped in the \ttt{PYINIT} call.
\iteme{= 1 :} yes, the \ttt{PYINIT} execution will finish normally,
but with the flag \ttt{MSTI(53) = 1} set to signal the problem. If
nevertheless \ttt{PYEVNT} is called after this, the run will be
stopped, since no events can be generated. If instead a new
\ttt{PYINIT} call is made, with changed conditions (e.g. modified
supersymmetry parameters in a SUSY run), it may now become possible
to initialize normally and generate events.
\end{subentry}
\iteme{MSTP(128) :} (D = 0) storing of copy of resonance decay
products in the documentation section of the event record, and
mother pointer (\ttt{K(I,3)}) relation of the actual resonance
decay products (stored in the main section of the event record)
to the documentation copy.
\begin{subentry}
\iteme{= 0 :} products are stored also in the documentation section,
and each product stored in the main section points back
to the corresponding entry in the documentation section.
\iteme{= 1 :} products are stored also in the documentation section,
but the products stored in the main section point back to
the decaying resonance copy in the main section.
\iteme{= 2 :} products are not stored in the documentation section;
the products stored in the main section point back to the
decaying resonance copy in the main section.
\end{subentry}
\iteme{MSTP(129) :} (D = 10) for the maximization of $2 \to 3$ processes
(\ttt{ISET(ISUB) = 5}) each phase-space point in $\tau$, $y$ and $\tau'$
is tested \ttt{MSTP(129)} times in the other dimensions (at randomly
selected points) to determine the effective maximum in the
($\tau$, $y$, $\tau'$) point.
\iteme{MSTP(131) :} (D = 0) master switch for pile-up events, i.e.\ several
independent hadron--hadron interactions generated in the same
bunch--bunch crossing, with the events following one after the
other in the event record. See section \ref{ss:pileup} for details.
\begin{subentry}
\iteme{= 0 :} off, i.e.\ only one event is generated at a time.
\iteme{= 1 :} on, i.e.\ several events are allowed in the same event
record. Information on the processes generated may be found in
\ttt{MSTI(41) - MSTI(50)}.
\end{subentry}
\iteme{MSTP(132) - MSTP(134) :} further switches for pile-up events,
see section \ref{ss:multintpar}.
\iteme{MSTP(141) :} (D = 0) calling of \ttt{PYKCUT} in the
event-generation chain, for inclusion of user-specified cuts.
\begin{subentry}
\iteme{= 0 :} not called.
\iteme{= 1 :} called.
\end{subentry}
\iteme{MSTP(142) :} (D = 0) calling of \ttt{PYEVWT} in the
event-generation chain, either to give weighted events or to modify
standard cross sections. See \ttt{PYEVWT} description in section
\ref{ss:PYTmainroutines} for further details.
\begin{subentry}
\iteme{= 0 :} not called.
\iteme{= 1 :} called; the distribution of events among subprocesses
and in kinematics variables is modified by the factor \ttt{WTXS},
set by you in the \ttt{PYEVWT} call, but events come with a
compensating weight \ttt{PARI(10) = 1./WTXS}, such that total
cross sections are unchanged.
\iteme{= 2 :} called; the cross section itself is modified by the
factor \ttt{WTXS}, set by you in the \ttt{PYEVWT} call.
\end{subentry}
\iteme{MSTP(143) :} (D = 0) calling of \ttt{UPVETO} in the
event-generation chain, to give the possibly to abort the generation
of an event.
\begin{subentry}
\iteme{= 0 :} not called, so no events are aborted (for this reason).
\iteme{= 1 :} yes, \ttt{UPVETO} is called, from inside the \ttt{PYEVNT}
routine (but not from \ttt{PYEVNW}), and a user can then decide whether
to abort the current event or not.
\end{subentry}
\iteme{MSTP(145) :} (D = 0) choice of polarization state for NRQCD
production of charmonium or bottomonium, processes in the ranges
421--439 and 461--479.
\begin{subentry}
\iteme{= 0 :} unpolarized squared partonic amplitude.
\iteme{= 1 :} helicity or density matrix elements, as chosen by
\ttt{MSTP(146)} and \ttt{MSTP(147)}. Only intended for experts.
\end{subentry}
\iteme{MSTP(146) :} (D = 1) choice of polarization reference frame when
\ttt{MSTP(145) = 1}.
\begin{subentry}
\iteme{= 1 :} recoil (recommended since it matches how {\Py} defines
particle directions, which the others do not obviously do).
\iteme{= 2 :} Gottfried--Jackson.
\iteme{= 3 :} target.
\iteme{= 4 :} Collins--Soper.
\end{subentry}
\iteme{MSTP(147) :} (D = 0) particular helicity or density matrix
component when \ttt{MSTP(145) = 1}.
\begin{subentry}
\iteme{= 0 :} helicity 0.
\iteme{= 1 :} helicity $\pm 1$.
\iteme{= 2 :} helicity $\pm 2$.
\iteme{= 3 :} density matrix element $\rho_{0,0}$.
\iteme{= 4 :} density matrix element $\rho_{1,1}$.
\iteme{= 5 :} density matrix element $\rho_{1,0}$.
\iteme{= 6 :} density matrix element $\rho_{1,-1}$.
\end{subentry}
\iteme{MSTP(148) :} (D = 0) possibility to allow final-state shower
evolution of the $\c\cbar[^3S_1^{(8)}]$ and $\b\bbar[^3S_1^{(8)}]$
states produced in the NRQCD production of charmonium or bottomonium.
Switching it on may exaggerate shower effects, since not all
$\Q\Qbar[^3S_1^{(8)}]$ comes from the fragmentation component where
radiation is expected.
\begin{subentry}
\iteme{= 0 :} off.
\iteme{= 1 :} on.
\end{subentry}
\iteme{MSTP(149) :} (D = 0) if the $\Q\Qbar[^3S_1^{(8)}]$ states are
allowed to radiate, \ttt{MSTP(148) = 1}, it determines the kinematics
of the $\Q\Qbar[^3S_1^{(8)}] \to \Q\Qbar[^3S_1^{(8)}] + \g$ branching.
\begin{subentry}
\iteme{= 0 :} always pick the $\Q\Qbar[^3S_1^{(8)}]$ to be the harder,
i.e.\ $z > 0.5$.
\iteme{= 1 :} allow $z < 0.5$ and $z > 0.5$ equally.
\end{subentry}
\iteme{MSTP(151) :} (D = 0) introduce smeared position of primary vertex
of events.
\begin{subentry}
\iteme{= 0 :} no, i.e.\ the primary vertex of each event is at the
origin.
\iteme{= 1 :} yes, with Gaussian distributions separately in $x$, $y$,
$z$ and $t$. The respective widths of the Gaussians have to be given
in \ttt{PARP(151) - PARP(154)}. Also pile-up events obtain separate
primary vertices. No provisions are made for more complicated
beam-spot shapes, e.g.\ with a spread in $z$ that varies as a
function of $t$. Note that a large beam spot combined with some of the
\ttt{MSTJ(22)} options may lead to many particles not being allowed to
decay at all.
\end{subentry}
\iteme{MSTP(161) :} (D = 0) unit number of file on which \ttt{PYUPIN}
should write its initialization info, and from which \ttt{UPINIT} should
read it back in, in cases where the Les Houches Accord is used to store
{\Py} hard processes.
\iteme{MSTP(162) :} (D = 0) unit number of file on which \ttt{PYUPEV}
should write its event info, and from which \ttt{UPEVNT} should
read it back in, in cases where the Les Houches Accord is used to store
{\Py} hard processes.
\iteme{MSTP(171) :} (D = 0) possibility of variable energies from one
event to the next. For further details see section \ref{ss:PYvaren}.
\begin{subentry}
\iteme{= 0 :} no; i.e.\ the energy is fixed at the initialization call.
\iteme{= 1 :} yes; i.e.\ a new energy has to be given for each new
event.
\itemc{Warning:} variable energies cannot be used in conjunction with
the internal generation of a virtual photon flux obtained by a
\ttt{PYINIT} call with {\galep} argument. The reason is that
a variable-energy machinery is now used internally for the $\gamma$-hadron
or $\gamma\gamma$ subsystem, with some information saved at
initialization for the full energy.
\end{subentry}
\iteme{MSTP(172) :} (D = 2) options for generation of events with
variable energies, applicable when \ttt{MSTP(171) = 1}.
\begin{subentry}
\iteme{= 1 :} an event is generated at the requested energy, i.e.\
internally a loop is performed over possible event configurations
until one is accepted. If the requested c.m.\ energy of an event
is below \ttt{PARP(2)} the run is aborted. Cross-section information
can not be trusted with this option, since it depends on how you
decided to pick the requested energies.
\iteme{= 2 :} only one event configuration is tried. If that is
accepted, the event is generated in full. If not, no event is
generated, and the status code \ttt{MSTI(61) = 1} is returned.
You are then expected to give a new energy, looping until an
acceptable event is found. No event is generated if the
requested c.m.\ energy is below \ttt{PARP(2)}, instead
\ttt{MSTI(61) = 1} is set to signal the failure. In principle,
cross sections should come out correctly with this option.
\end{subentry}
\iteme{MSTP(173) :} (D = 0) possibility for you to give in an event
weight to compensate for a biased choice of beam spectrum.
\begin{subentry}
\iteme{= 0 :} no, i.e.\ event weight is unity.
\iteme{= 1 :} yes; weight to be given for each event in
\ttt{PARP(173)}, with maximum weight given at initialization
in \ttt{PARP(174)}.
\end{subentry}
\iteme{MSTP(181) :} (R) {\Py} version number.
\iteme{MSTP(182) :} (R) {\Py} subversion number.
\iteme{MSTP(183) :} (R) last year of change for {\Py}.
\iteme{MSTP(184) :} (R) last month of change for {\Py}.
\iteme{MSTP(185) :} (R) last day of change for {\Py}.
\boxsep
\iteme{PARP(1) :}\label{p:PARP} (D = 0.25 GeV) nominal
$\Lambda_{\mrm{QCD}}$ used in running $\alphas$ for hard
scattering (see \ttt{MSTP(3)}).
\iteme{PARP(2) :} (D = 10. GeV) lowest c.m.\ energy for the
event as a whole that the program will accept to simulate.
\iteme{PARP(13) :} (D = 1. GeV$^2$) $Q_{\mmax}^2$ scale, to be set by
you for defining maximum scale allowed for photoproduction when
using the option \ttt{MSTP(13) = 2}.
\iteme{PARP(14) :} (D = 0.01) in the numerical integration of quark
and gluon parton distributions inside an electron, the successive
halvings of evaluation-point spacing is interrupted when two values
agree in relative size, $|$new$-$old$|$/(new$+$old), to better than
\ttt{PARP(14)}. There are hardwired lower and upper limits of 2 and
8 halvings, respectively.
\iteme{PARP(15) :} (D = 0.5 GeV) lower cut-off $p_0$ used to define
minimum transverse momentum in branchings $\gamma \to \q\qbar$ in
the anomalous event class of $\gamma\p$ interactions, i.e.\ sets the
dividing line between the VMD and GVMD event classes.
\iteme{PARP(16) :} (D = 1.) the anomalous parton-distribution functions
of the photon are taken to have the charm and bottom flavour thresholds
at virtuality \ttt{PARP(16)}$\times m_{\q}^2$.
\iteme{PARP(17) :} (D = 1.) rescaling factor used for the $Q$ argument
of the anomalous parton distributions of the photon, see
\ttt{MSTP(15)}.
\iteme{PARP(18) :} (D = 0.4 GeV) scale $k_{\rho}$, such that the cross
sections of a GVMD state of scale $\kT$ is suppressed by a factor
$k_{\rho}^2/\kT^2$ relative to those of a VMD state. Should be of
order $m_{\rho}/2$, with some finetuning to fit data.
\iteme{PARP(25) :} (D = 0.) parameter $\eta$ describing the admixture
of CP-odd Higgs decays for \ttt{MSTP(25) = 3}.
\iteme{PARP(31) :} (D = 1.5) common $K$ factor multiplying the
differential cross section for hard parton--parton processes
when \ttt{MSTP(33) = 1} or \ttt{2}, with the exception of colour
annihilation graphs in the latter case.
\iteme{PARP(32) :} (D = 2.0) special $K$ factor multiplying the
differential cross section in hard colour annihilation graphs,
including resonance production, when \ttt{MSTP(33) = 2}.
\iteme{PARP(33) :} (D = 0.075) this factor is used to multiply the
ordinary $Q^2$ scale in $\alphas$ at the hard interaction for
\ttt{MSTP(33) = 3}. With the default value, which is only to be taken as
an example, the effective $K$ factor thus obtained for jet production
is in accordance with the NLO results in \cite{Ell86}, modulo the
danger of double-counting because of parton-shower corrections to
jet rates.
\iteme{PARP(34) :} (D = 1.) the $Q^2$ scale defined by \ttt{MSTP(32)} is
multiplied by \ttt{PARP(34)} when it is used as argument for parton
distributions and $\alphas$ at the hard interaction. It does not affect
$\alphas$ when \ttt{MSTP(33) = 3}, nor does it change the $Q^2$ argument
of parton showers.
\iteme{PARP(35) :} (D = 0.20) fix $\alphas$ value that is used in
the heavy-flavour threshold factor when \ttt{MSTP(35) = 1}.
\iteme{PARP(36) :} (D = 0. GeV) the width $\Gamma_{\Q}$ for the heavy
flavour studied in processes {\ISUB} = 81 or 82; to be used for the
threshold factor when \ttt{MSTP(35) = 2}.
\iteme{PARP(37) :} (D = 1.) for \ttt{MSTP(37) = 1} this regulates the
point at which the reference on-shell quark mass in Higgs and
technicolor couplings is assumed defined in \ttt{PYMRUN} calls;
specifically the running quark mass is assumed to coincide with the
fix one at an energy scale \ttt{PARP(37)} times the fix quark mass,
i.e.\ $m_{\mrm{running}}($\ttt{PARP(37)}$\times m_{\mrm{fix}}) =
m_{\mrm{fix}}$. See discussion at eq.~(\ref{eq:runqmass}) on ambiguity
of \ttt{PARP(37)} choice.
\iteme{PARP(38) :} (D = 0.70 GeV$^3$) the squared wave function at the
origin, $|R(0)|^2$, of the $\Jpsi$ wave function. Used for processes
86 and 106--108. See ref. \cite{Glo88}.
\iteme{PARP(39) :} (D = 0.006 GeV$^3$) the squared derivative of the
wave function at the origin, $|R'(0)|^2/m^2$, of the $\chi_{\c}$
wave functions. Used for processes 87--89 and 104--105. See ref.
\cite{Glo88}.
\iteme{PARP(41) :} (D = 0.020 GeV) in the process of generating mass
for resonances, and optionally to force that mass to be in a given
range, only resonances with a total width in excess of \ttt{PARP(41)}
are generated according to a Breit--Wigner shape (if allowed by
\ttt{MSTP(42)}), while narrower resonances are put on the mass
shell.
\iteme{PARP(42) :} (D = 2. GeV) minimum mass of resonances assumed
to be allowed when evaluating total width of $\hrm^0$ to $\Z^0 \Z^0$ or
$\W^+ \W^-$ for cases when the $\hrm^0$ is so light that (at least)
one $\Z/\W$ is forced to be off the mass shell. Also generally used
as safety check on minimum mass of resonance. Note that some
\ttt{CKIN} values may provide additional constraints.
\iteme{PARP(43) :} (D = 0.10) precision parameter used in numerical
integration of width for a channel with at least one daughter off
the mass shell.
\iteme{PARP(44) :} (D = 1000.) the $\nu$ parameter of the strongly
interacting $\Z/\W$ model of Dobado, Herrero and Terron \cite{Dob91};
see \ttt{MSTP(46) = 3}.
\iteme{PARP(45) :} (D = 2054. GeV) the effective techni-$\rho$ mass
parameter of the strongly interacting model of Dobado, Herrero and
Terron \cite{Dob91}; see \ttt{MSTP(46) = 5}. On physical grounds it
should not be chosen smaller than about 1 TeV or larger than
about the default value.
\iteme{PARP(46) :} (D = 123. GeV) the $F_{\pi}$ decay constant that
appears inversely quadratically in all techni-$\eta$ partial decay
widths \cite{Eic84,App92}.
\iteme{PARP(47) :} (D = 246. GeV) vacuum expectation value $v$ used
in the DHT scenario \cite{Dob91} to define the width of the
techni-$\rho$; this width is inversely proportional $v^2$.
\iteme{PARP(48) :} (D = 50.) the Breit--Wigner factor in the cross section
is set to vanish for masses that deviate from the nominal one by
more than \ttt{PARP(48)} times the nominal resonance width (i.e.\ the
width evaluated at the nominal mass). Is used in most processes
with a single $s$-channel resonance, but there are some exceptions,
notably $\gamma^*/\Z^0$ and $\W^{\pm}$. The reason for this option
is that the conventional Breit--Wigner description is at times not
really valid far away from the resonance position, e.g.\ because of
interference with other graphs that should then be included. The wings
of the Breit--Wigner can therefore be removed.
\iteme{PARP(50) :} (D = 0.054) dimensionless coupling, which enters
quadratically in all partial widths of the excited graviton $\G^*$
resonance, is
$\kappa m_{\G^*} = \sqrt{2} x_1 k /\br{M}_{\mrm{Pl}}$,
where $x_1 \approx 3.83$ is the first zero of the $J_1$ Bessel function
and $\br{M}_{\mrm{Pl}}$ is the modified Planck mass scale
\cite{Ran99,Bij01}.
\iteme{PARP(61) - PARP(65) :} (C) parameters for initial-state
radiation, see section \ref{ss:showrout}.
\iteme{PARP(71) - PARP(72) :} (C) parameter for final-state
radiation, see section \ref{ss:showrout}.
\iteme{PARP(78) - PARP(90) :} parameters for multiple interactions,
see section \ref{ss:multintpar}.
\iteme{PARP(91) - PARP(100) :} parameters for beam-remnant
treatment, see section \ref{ss:multintpar}.
\iteme{PARP(101) :} (D = 0.50) fraction of diffractive systems in which
a quark is assumed kicked out by the pomeron rather than a gluon;
applicable for option \ttt{MSTP(101) = 3}.
\iteme{PARP(102) :} (D = 0.28 GeV) the mass spectrum of diffractive
states (in single and double diffractive scattering) is assumed to
start \ttt{PARP(102)} above the mass of the particle that is
diffractively excited. In this connection, an incoming $\gamma$
is taken to have the selected VMD meson mass, i.e.\ $m_{\rho}$,
$m_{\omega}$, $m_{\phi}$ or $m_{\Jpsi}$.
\iteme{PARP(103) :} (D = 1.0 GeV) if the mass of a diffractive state
is less than \ttt{PARP(103)} above the mass of the particle that is
diffractively excited, the state is forced to decay isotropically
into a two-body channel. In this connection, an incoming $\gamma$
is taken to have the selected VMD meson mass, i.e.\ $m_{\rho}$,
$m_{\omega}$, $m_{\phi}$ or $m_{\Jpsi}$. If the mass is higher than
this threshold, the standard string fragmentation machinery is used.
The forced two-body decay is always carried out, also when
\ttt{MSTP(111) = 0}.
\iteme{PARP(104) :} (D = 0.8 GeV) minimum energy above threshold for
which hadron--hadron total, elastic and diffractive cross sections
are defined. Below this energy, an alternative description in terms
of specific few-body channels would have been required, and this
is not modelled in {\Py}.
\iteme{PARP(110) :} (D = 1.) a rescaling factor for resonance widths,
applied when \ttt{MSTP(110)} is switched on.
\iteme{PARP(111) :} (D = 2. GeV) used to define the minimum invariant
mass of the remnant hadronic system (i.e.\ when interacting partons
have been taken away), together with original hadron masses and
extra parton masses. For a hadron or resolved photon beam, this also
implies a further constraint that the $x$ of an interacting parton
be below $1 - 2 \times \mbox{\ttt{PARP(111)}}/E_{\mrm{cm}}$.
\iteme{PARP(115) :} (D = 1.5 fm) (C) the average fragmentation time
of a string, giving the exponential suppression that a reconnection
cannot occur if strings decayed before crossing. Is implicitly
fixed by the string constant and the fragmentation function
parameters, and so a significant change is not recommended.
\iteme{PARP(116) :} (D = 0.5 fm) (C) width of the type I string
in reconnection calculations, giving the radius of the Gaussian
distribution in $x$ and $y$ separately.
\iteme{PARP(117) :} (D = 0.6) (C) $k_{\mrm{I}}$, the main free parameter
in the reconnection probability for scenario I; the probability is
given by \ttt{PARP(117)} times the overlap volume, up to saturation
effects.
\iteme{PARP(118), PARP(119) :} (D = 2.5, 2.0) (C) $f_r$ and $f_t$,
respectively, used in the Monte Carlo sampling of the phase space volume
in scenario I. There is no real reason to change these numbers.
\iteme{PARP(120) :} (D = 1.0) (D) (C) fraction of events in the GH,
intermediate and instantaneous scenarios where a reconnection is allowed
to occur. For the GH one a further suppression of the reconnection
rate occurs from the requirement of reduced string length in a
reconnection.
\iteme{PARP(121) :} (D = 1.) the maxima obtained at initial
maximization are multiplied by this factor if \ttt{MSTP(121) = 1};
typically \ttt{PARP(121)} would be given as the product of the
violation factors observed (i.e.\ the ratio of final maximum value
to initial maximum value) for the given process(es).
\iteme{PARP(122) :} (D = 0.4) fraction of total probability that is
shared democratically between the \ttt{COEF} coefficients open for
the given variable, with the remaining fraction distributed according
to the optimization results of \ttt{PYMAXI}.
\iteme{PARP(131) :} parameter for pile-up events, see section
\ref{ss:multintpar}.
\iteme{PARP(141) - PARP(150) :} (D = 10*1.) matrix elements for
charmonium and bottomonium production in the non-relativistic
QCD framework (NRQCD). Current values are dummy only, and will be
updated soon. These values are used in processes 421--439 and
461--479.
\begin{subentry}
\iteme{PARP(141) :} $\langle \mathcal{O}^{\Jpsi}[^3S_1^{(1)}] \rangle$.
\iteme{PARP(142) :} $\langle \mathcal{O}^{\Jpsi}[^3S_1^{(8)}] \rangle$.
\iteme{PARP(143) :} $\langle \mathcal{O}^{\Jpsi}[^1S_0^{(8)}] \rangle$.
\iteme{PARP(144) :} $\langle \mathcal{O}^{\Jpsi}[^3P_0^{(8)}] %
\rangle / m_{\c}^2$.
\iteme{PARP(145) :} $\langle \mathcal{O}^{\chi_{\c 0}}[^3P_0^{(1)}] %
\rangle / m_{\c}^2$.
\iteme{PARP(146) :} $\langle \mathcal{O}^{\Upsilon}[^3S_1^{(1)}] \rangle$.
\iteme{PARP(147) :} $\langle \mathcal{O}^{\Upsilon}[^3S_1^{(8)}] \rangle$.
\iteme{PARP(148) :} $\langle \mathcal{O}^{\Upsilon}[^1S_0^{(8)}] \rangle$.
\iteme{PARP(149) :} $\langle \mathcal{O}^{\Upsilon}[^3P_0^{(8)}] %
\rangle / m_{\b}^2$.
\iteme{PARP(150) :} $\langle \mathcal{O}^{\chi_{\b 0}}[^3P_0^{(1)}] %
\rangle / m_{\b}^2$.
\end{subentry}
\iteme{PARP(151) - PARP(154) :} (D = 4*0.) (C) regulate the assumed
beam-spot size. For \ttt{MSTP(151) = 1} the $x$, $y$, $z$ and $t$
coordinates of the primary vertex of each event are selected
according to four independent Gaussians. The widths of these
Gaussians are given by the four parameters, where the first three
are in units of mm and the fourth in mm/$c$.
\iteme{PARP(161) - PARP(164) :} (D = 2.20, 23.6, 18.4, 11.5) couplings
$f_V^2/4\pi$ of the photon to the $\rho^0$, $\omega$, $\phi$ and
$\Jpsi$ vector mesons.
\iteme{PARP(165) :} (D = 0.5) a simple multiplicative factor applied to
the cross section for the transverse resolved photons to take into
account the effects of longitudinal resolved photons, see
\ttt{MSTP(17)}. No preferred value, but typically one could use
\ttt{PARP(165) = 1} as main contrast to the no-effect \ttt{= 0}, with the
default arbitrarily chosen in the middle.
\iteme{PARP(167), PARP(168) :} (D = 2*0) the longitudinal energy fraction
$y$ of an incoming photon, side 1 or 2, used in the $R$ expression
given for \ttt{MSTP(17)} to evaluate $f_L(y,Q^2)/f_T(y,Q^2)$. Need not
be supplied when a photon spectrum is generated inside a lepton beam,
but only when a photon is directly given as argument in the \ttt{PYINIT}
call.
\iteme{PARP(171) :} to be set, event-by-event, when variable
energies are allowed, i.e.\ when \ttt{MSTP(171) = 1}. If \ttt{PYINIT} is
called with \ttt{FRAME = 'CMS'} (\ttt{= 'FIXT'}), \ttt{PARP(171)}
multiplies the c.m.\ energy (beam energy) used at initialization.
For the options \ttt{'3MOM'}, \ttt{'4MOM'} and \ttt{'5MOM'},
\ttt{PARP(171)} is dummy, since there the momenta are set in the
\ttt{P} array. It is also dummy for the \ttt{'USER'} option,
where the choice of variable energies is beyond the control of {\Py}.
\iteme{PARP(173) :} event weight to be given by you when
\ttt{MSTP(173) = 1}.
\iteme{PARP(174) :} (D = 1.) maximum event weight that will be
encountered in \ttt{PARP(173)} during the course of a run with
\ttt{MSTP(173) = 1}; to be used to optimize the efficiency of the
event generation. It is always allowed to use a larger bound than
the true one, but with a corresponding loss in efficiency.
\iteme{PARP(181) - PARP(189) :} (D = 0.1, 0.01, 0.01, 0.01, 0.1, 0.01,
0.01, 0.01, 0.3) Yukawa couplings of leptons to $\H^{++}$, assumed
same for $\H_L^{++}$ and $\H_R^{++}$. Is a symmetric $3 \times 3$
array, where \ttt{PARP(177+3*i+j)} gives the coupling to a lepton pair
with generation indices $i$ and $j$. Thus the default matrix is
dominated by the diagonal elements and especially by the $\tau\tau$ one.
\iteme{PARP(190) :} (D = 0.64) $g_L = e/\sin\theta_W$.
\iteme{PARP(191) :} (D = 0.64) $g_R$, assumed same as $g_L$.
\iteme{PARP(192) :} (D = 5 GeV) $v_L$ vacuum expectation value of the
left-triplet. The corresponding $v_R$ is assumed given by
$v_R = \sqrt{2} M_{\W_R} / g_R$ and is not stored explicitly.
\iteme{PARP(193) :} (D = 1D4 GeV$^2$) factorization scale $Q^2$ for
parton densities, to be set by user when \ttt{MSTP(32) = 12} for
$2 \to 2$ processes or \ttt{MSTP(39) = 8} for $2 \to 3$ ones.
\iteme{PARP(194) :} (D = 1D4 GeV$^2$) renormalization scale $Q^2$,
to be set by user when \ttt{MSTP(32) = 12} for $2 \to 2$ processes
or \ttt{MSTP(39) = 8} for $2 \to 3$ ones. For process 161 it also sets
the scale of running quark masses.
\end{entry}
\subsection{Further Couplings}
\label{ss:coupcons}
In this section we collect information on the two routines for
running $\alphas$ and $\alphaem$, and on other couplings
of standard and non-standard particles found in the \ttt{PYDAT1} and
\ttt{PYTCSM} common blocks. Although originally begun for applications
within the traditional particle sector, this section of \ttt{PYDAT1}
has rapidly expanded towards the non-standard aspects, and is thus more
of interest for applications to specific processes. It could therefore
equally well have been put somewhere else in this manual. Several other
couplings indeed appear in the \ttt{PARP} array in the \ttt{PYPARS}
common block, see section \ref{ss:PYswitchpar}, and the choice between
the two has largely been dictated by availability of space. The
improved simulation of the TechniColor Strawman Model, described
in \cite{Lan02,Lan02a}, and the resulting proliferation of model
parameters, has led to the introduction of the new \ttt{PYTCSM}
common block.
\drawbox{ALEM = PYALEM(Q2)}\label{p:PYALEM}
\begin{entry}
\itemc{Purpose:} to calculate the running electromagnetic coupling
constant $\alphaem$. Expressions used are described in
ref. \cite{Kle89}. See \ttt{MSTU(101)}, \ttt{PARU(101)},
\ttt{PARU(103)} and \ttt{PARU(104)}.
\iteme{Q2 :} the momentum transfer scale $Q^2$ at which to evaluate
$\alphaem$.
\end{entry}
\drawbox{ALPS = PYALPS(Q2)}\label{p:PYALPS}
\begin{entry}
\itemc{Purpose:} to calculate the running strong coupling constant
$\alphas$, e.g.\ in matrix elements and resonance decay widths.
(The function is not used in parton showers, however, where
formulae rather are written in terms of the relevant $\Lambda$
values.) The first- and second-order expressions are given by
eqs.~(\ref{ee:aS3j}) and (\ref{ee:aS4j}). See
\ttt{MSTU(111) - MSTU(118)} and \ttt{PARU(111) - PARU(118)} for options.
\iteme{Q2 :} the momentum transfer scale $Q^2$ at which to evaluate
$\alphas$.
\end{entry}
\drawbox{PM = PYMRUN(KF,Q2)}\label{p:PYMRUN}
\begin{entry}
\itemc{Purpose:} to give running masses of $\d$, $\u$, $\s$, $\c$, $\b$
and $\t$ quarks according to eq.~(\ref{eq:runqmass}). For all other
particles, the \ttt{PYMASS} function is called by \ttt{PYMRUN} to give
the normal mass. Such running masses appear e.g.\ in couplings of
fermions to Higgs and technipion states.
\iteme{KF :} flavour code.
\iteme{Q2 :} the momentum transfer scale $Q^2$ at which to evaluate
$\alphas$.
\itemc{Note:} the nominal values, valid at a reference scale \\
$Q^2_{\mrm{ref}} = \max((\mtt{PARP(37)} m_{\mrm{nominal}})^2 , 4\Lambda^2)$,\\
are stored in \ttt{PARF(91) - PARF(96)}.
\end{entry}
\drawbox{COMMON/PYDAT1/MSTU(200),PARU(200),MSTJ(200),PARJ(200)}
\begin{entry}
\itemc{Purpose:} to give access to a number of status codes and
parameters which regulate the performance of the program as a whole.
Here only those related to couplings are described; the main
description is found in section \ref{ss:JETswitch}.
\boxsep
\iteme{MSTU(101) :}\label{p:MSTU101} (D = 1) procedure for
$\alphaem$ evaluation in the \ttt{PYALEM} function.
\begin{subentry}
\iteme{= 0 :} $\alphaem$ is taken fixed at the value
\ttt{PARU(101)}.
\iteme{= 1 :} $\alphaem$ is running with the $Q^2$ scale,
taking into account corrections from fermion loops ($\e$, $\mu$,
$\tau$, $\d$, $\u$, $\s$, $\c$, $\b$).
\iteme{= 2 :} $\alphaem$ is fixed, but with separate values at low
and high $Q^2$. For $Q^2$ below (above) \ttt{PARU(104)} the value
\ttt{PARU(101)} (\ttt{PARU(103)}) is used. The former value is
then intended for real photon emission, the latter for
electroweak physics, e.g.\ of the $\W/\Z$ gauge bosons.
\end{subentry}
\iteme{MSTU(111) :} (I, D=1) order of $\alphas$ evaluation in the
\ttt{PYALPS} function. Is overwritten in \ttt{PYEEVT}, \ttt{PYONIA} or
\ttt{PYINIT} calls with the value desired for the process under study.
\begin{subentry}
\iteme{= 0 :} $\alphas$ is fixed at the value \ttt{PARU(111)}.
As extra safety, $\Lambda=$\ttt{PARU(117)} is set in \ttt{PYALPS} so
that the first-order running $\alphas$ agrees with the desired fixed
$\alphas$ for the $Q^2$ value used.
\iteme{= 1 :} first-order running $\alphas$ is used.
\iteme{= 2 :} second-order running $\alphas$ is used.
\end{subentry}
\iteme{MSTU(112) :} (D = 5) the nominal number of flavours assumed in
the $\alphas$ expression, with respect to which $\Lambda$ is defined.
\iteme{MSTU(113) :} (D = 3) minimum number of flavours that may be
assumed in $\alphas$ expression, see \ttt{MSTU(112)}.
\iteme{MSTU(114) :} (D = 5) maximum number of flavours that may be
assumed in $\alphas$ expression, see \ttt{MSTU(112)}.
\iteme{MSTU(115) :} (D = 0) treatment of $\alphas$ singularity for
$Q^2 \to 0$ in \ttt{PYALPS} calls. (Relevant e.g. for QCD $2 \to 2$
matrix elements in the $\pT \to 0$ limit, but not for showers, where
\ttt{PYALPS} is not called.)
\begin{subentry}
\iteme{= 0 :} allow it to diverge like $1/\ln(Q^2/\Lambda^2)$.
\iteme{= 1 :} soften the divergence to $1/\ln(1 + Q^2/\Lambda^2)$.
\iteme{= 2 :} freeze $Q^2$ evolution below \ttt{PARU(114)}, i.e.\
the effective argument is $\max(Q^2, $\ttt{PARU(114)}$)$.
\end{subentry}
\iteme{MSTU(118) :} (I) number of flavours $n_f$ found and used in
latest \ttt{PYALPS} call.
\boxsep
\iteme{PARU(101) :}\label{p:PARU101} (D = 0.00729735=1/137.04)
$\alphaem$, the electromagnetic fine structure constant at
vanishing momentum transfer.
\iteme{PARU(102) :} (D = 0.232) $\ssintw$, the weak mixing angle of the
standard electroweak model.
\iteme{PARU(103) :} (D = 0.007764=1/128.8) typical $\alphaem$ in
electroweak processes; used for $Q^2 >$ \ttt{PARU(104)} in the
option \ttt{MSTU(101) = 2} of \ttt{PYALEM}. Although it can technically
be used also at rather small $Q^2$, this $\alphaem$ value is mainly
intended for high $Q^2$, primarily $\Z^0$ and $\W^{\pm}$ physics.
\iteme{PARU(104) :} (D = 1 GeV$^2$) dividing line between `low' and
`high' $Q^2$ values in the option \ttt{MSTU(101) = 2} of \ttt{PYALEM}.
\iteme{PARU(105) :} (D = 1.16639E-5 GeV$^{-2}$) $G_{\F}$, the Fermi
constant of weak interactions.
\iteme{PARU(108) :} (I) the $\alphaem$ value obtained in the
latest call to the \ttt{PYALEM} function.
\iteme{PARU(111) :} (D = 0.20) fix $\alphas$ value assumed in
\ttt{PYALPS} when \ttt{MSTU(111) = 0} (and also in parton showers
when $\alphas$ is assumed fix there).
\iteme{PARU(112) :} (I, D=0.25 GeV) $\Lambda$ used in running $\alphas$
expression in \ttt{PYALPS}. Like \ttt{MSTU(111)}, this value is
overwritten by the calling physics routines, and is therefore purely
nominal.
\iteme{PARU(113) :} (D = 1.) the flavour thresholds, for the effective
number of flavours $n_f$ to use in the $\alphas$ expression, are
assumed to sit at $Q^2 = $\ttt{PARU(113)}$\times m_{\q}^2$, where
$m_{\q}$ is the quark mass. May be overwritten from the calling
physics routine.
\iteme{PARU(114) :} (D = 4 GeV$^2$) $Q^2$ value below which the
$\alphas$ value is assumed constant for \ttt{MSTU(115) = 2}.
\iteme{PARU(115) :} (D = 10.) maximum $\alphas$ value that \ttt{PYALPS}
will ever return; is used as a last resort to avoid singularities.
\iteme{PARU(117) :} (I) $\Lambda$ value (associated with
\ttt{MSTU(118)} effective flavours) obtained in latest \ttt{PYALPS}
call.
\iteme{PARU(118) :} (I) $\alphas$ value obtained in latest
\ttt{PYALPS} call.
\iteme{PARU(121) - PARU(130) :} couplings of a new $\Z'^0$; for
fermion default values are given by the Standard Model $\Z^0$ values,
assuming $\ssintw = 0.23$. Since a generation dependence is now
allowed for the $\Z'^0$ couplings to fermions, the variables
\ttt{PARU(121) - PARU(128)} only refer to the first generation,
with the second generation in \ttt{PARJ(180) - PARJ(187)} and
the third in \ttt{PARJ(188) - PARJ(195)} following exactly the
same pattern. Note that e.g.\ the $\Z'^0$ width contains
squared couplings, and thus depends quadratically on the values below.
\begin{subentry}
\iteme{PARU(121), PARU(122) :} (D = $-0.693$, $-1.$) vector and axial
couplings of down type quarks to $\Z'^0$.
\iteme{PARU(123), PARU(124) :} (D = 0.387, 1.) vector and axial
couplings of up type quarks to $\Z'^0$.
\iteme{PARU(125), PARU(126) :} (D = $-0.08$, $-1.$) vector and axial
couplings of leptons to $\Z'^0$.
\iteme{PARU(127), PARU(128) :} (D = 1., 1.) vector and axial
couplings of neutrinos to $\Z'^0$.
\iteme{PARU(129) :} (D = 1.) the coupling $Z'^0 \to \W^+ \W^-$ is
taken to be \ttt{PARU(129)}$\times$(the Standard Model
$\Z^0 \to \W^+ \W^-$ coupling)$\times (m_{\W}/m_{\Z'})^2$.
This gives a $\Z'^0 \to \W^+ \W^-$ partial width that
increases proportionately to the $\Z'^0$ mass.
\iteme{PARU(130) :} (D = 0.) in the decay chain
$\Z'^0 \to \W^+ \W^- \to 4$ fermions, the angular distribution in
the $\W$ decays is supposed to be a mixture, with fraction
\ttt{1. - PARU(130)} corresponding to the same angular distribution
between the four final fermions as in $\Z^0 \to \W^+ \W^-$ (mixture
of transverse and longitudinal $\W$'s), and fraction \ttt{PARU(130)}
corresponding to $\hrm^0 \to \W^+ \W^-$ the same way (longitudinal
$\W$'s).
\end{subentry}
\iteme{PARU(131) - PARU(136) :} couplings of a new $\W'^{\pm}$;
for fermions default values are given by the Standard Model
$\W^{\pm}$ values (i.e.\ $V-A$). Note that e.g.\ the $\W'^{\pm}$
width contains squared couplings, and
thus depends quadratically on the values below.
\begin{subentry}
\iteme{PARU(131), PARU(132) :} (D = 1., $-1.$) vector and axial couplings
of a quark--antiquark pair to $\W'^{\pm}$; is further multiplied by the
ordinary CKM factors.
\iteme{PARU(133), PARU(134) :} (D = 1., $-1.$) vector and axial couplings
of a lepton-neutrino pair to $\W'^{\pm}$.
\iteme{PARU(135) :} (D = 1.) the coupling $\W'^{\pm} \to \Z^0 \W^{\pm}$
is taken to be \ttt{PARU(135)}$\times$(the Standard Model
$\W^{\pm} \to \Z^0 \W^{\pm}$ coupling)$\times (m_{\W}/m_{W'})^2$.
This gives a $\W'^{\pm} \to \Z^0 \W^{\pm}$ partial width that
increases proportionately to the $\W'$ mass.
\iteme{PARU(136) :} (D = 0.) in the decay chain
$\W'^{\pm} \to \Z^0 \W^{\pm} \to 4$ fermions,
the angular distribution in the $\W/\Z$ decays is supposed to be a
mixture, with fraction \ttt{1-PARU(136)} corresponding to the same
angular distribution between the four final fermions as in
$\W^{\pm} \to \Z^0 \W^{\pm}$ (mixture of transverse and longitudinal
$\W/\Z$'s), and fraction \ttt{PARU(136)} corresponding to
$\H^{\pm} \to \Z^0 \W^{\pm}$ the same way (longitudinal $\W/\Z$'s).
\end{subentry}
\iteme{PARU(141) :} (D = 5.) $\tan\beta$ parameter of a two Higgs
doublet scenario, i.e.\ the ratio of vacuum expectation values.
This affects mass relations and couplings in the Higgs sector.
If the Supersymmetry simulation is switched on, \ttt{IMSS(1)}
nonvanishing, \ttt{PARU(141)} will be overwritten by \ttt{RMSS(5)}
at initialization, so it is the latter variable that should be set.
\iteme{PARU(142) :} (D = 1.) the $\Z^0 \to \H^+ \H^-$ coupling is
taken to be \ttt{PARU(142)}$\times$(the MSSM $\Z^0 \to \H^+ \H^-$
coupling).
\iteme{PARU(143) :} (D = 1.) the $\Z'^0 \to \H^+ \H^-$ coupling is
taken to be \ttt{PARU(143)}$\times$(the MSSM $\Z^0 \to \H^+ \H^-$
coupling).
\iteme{PARU(145) :} (D = 1.) quadratically multiplicative factor in the
$\Z'^0 \to \Z^0 \hrm^0$ partial width in left--right-symmetric models,
expected to be unity (see \cite{Coc91}).
\iteme{PARU(146) :} (D = 1.) $\sin(2\alpha)$ parameter, enters
quadratically as multiplicative factor in the
$\W'^{\pm} \to \W^{\pm} \hrm^0$ partial width in
left--right-symmetric models (see \cite{Coc91}).
\iteme{PARU(151) :} (D = 1.) multiplicative factor in the
$\L_{\Q} \to \q \ell$ squared Yukawa coupling, and thereby in the
$\L_{\Q}$ partial width and the $\q \ell \to \L_{\Q}$ and other
cross sections. Specifically,
$\lambda^2/(4\pi) = $\ttt{PARU(151)}$\times \alphaem$, i.e.\ it
corresponds to the $k$ factor of \cite{Hew88}.
\iteme{PARU(161) - PARU(168) :} (D = 5*1., 3*0.) multiplicative factors
that can be used to modify the default couplings of the $\hrm^0$
particle in {\Py}. Note that the factors enter quadratically in the
partial widths. The default values correspond to the couplings given
in the minimal one-Higgs-doublet Standard Model, and are therefore
not realistic in a two-Higgs-doublet scenario. The default values
should be changed appropriately by you. Also the last two default
values should be changed; for these the expressions of the
minimal supersymmetric Standard Model (MSSM) are given to show
parameter normalization. Alternatively, the SUSY machinery can
generate all the couplings for \ttt{IMSS(1)}, see \ttt{MSTP(4)}.
\begin{subentry}
\iteme{PARU(161) :} $\hrm^0$ coupling to down type quarks.
\iteme{PARU(162) :} $\hrm^0$ coupling to up type quarks.
\iteme{PARU(163) :} $\hrm^0$ coupling to leptons.
\iteme{PARU(164) :} $\hrm^0$ coupling to $\Z^0$.
\iteme{PARU(165) :} $\hrm^0$ coupling to $\W^{\pm}$.
\iteme{PARU(168) :} $\hrm^0$ coupling to $\H^{\pm}$ in
$\gamma\gamma \to \hrm^0$ loops, in MSSM
$\sin(\beta-\alpha)+\cos(2\beta)\sin(\beta+\alpha) /
(2\scostw)$.
\end{subentry}
\iteme{PARU(171) - PARU(178) :} (D = 7*1., 0.) multiplicative factors
that can be used to modify the default couplings of the $\H^0$
particle in {\Py}. Note that the factors enter quadratically in
partial widths. The default values for \ttt{PARU(171) - PARU(175)}
correspond to the couplings given to $\hrm^0$ in the minimal
one-Higgs-doublet Standard Model, and are therefore not realistic
in a two-Higgs-doublet scenario. The default values should
be changed appropriately by you. Also the last two default
values should be changed; for these the expressions of the
minimal supersymmetric Standard Model (MSSM) are given to show
parameter normalization. Alternatively, the SUSY machinery can generate
all the couplings for \ttt{IMSS(1)}, see \ttt{MSTP(4)}.
\begin{subentry}
\iteme{PARU(171) :} $\H^0$ coupling to down type quarks.
\iteme{PARU(172) :} $\H^0$ coupling to up type quarks.
\iteme{PARU(173) :} $\H^0$ coupling to leptons.
\iteme{PARU(174) :} $\H^0$ coupling to $\Z^0$.
\iteme{PARU(175) :} $\H^0$ coupling to $W^{\pm}$.
\iteme{PARU(176) :} $\H^0$ coupling to $\hrm^0 \hrm^0$, in MSSM
$\cos(2\alpha) \cos(\beta+\alpha) - 2 \sin(2\alpha)
\sin(\beta+\alpha)$.
\iteme{PARU(177) :} $\H^0$ coupling to $\A^0 \A^0$, in MSSM
$\cos(2\beta) \cos(\beta+\alpha)$.
\iteme{PARU(178) :} $\H^0$ coupling to $\H^{\pm}$ in
$\gamma \gamma \to \H^0$ loops, in MSSM
$\cos(\beta-\alpha) - \cos(2\beta)\cos(\beta+\alpha) /
(2\scostw)$.
\end{subentry}
\iteme{PARU(181) - PARU(190) :} (D = 3*1., 2*0., 2*1., 3*0.)
multiplicative factors that can be used to modify the default
couplings of the $\A^0$ particle in {\Py}. Note that the factors
enter quadratically in partial widths. The default values for
\ttt{PARU(181) - PARU(183)} correspond
to the couplings given to $\hrm^0$ in the minimal one-Higgs-doublet
Standard Model, and are therefore not realistic in a
two-Higgs-doublet scenario. The default values should
be changed appropriately by you. \ttt{PARU(184)} and \ttt{PARU(185)}
should be vanishing at the tree level, in the absence of
CP-violating phases in the Higgs sector, and are so set;
normalization of these couplings agrees with what is used for
$\hrm^0$ and $\H^0$. Also the other default values should be changed; for
these the expressions of the Minimal Supersymmetric Standard
Model (MSSM) are given to show parameter normalization. Alternatively,
the SUSY machinery can generate all the couplings for \ttt{IMSS(1)},
see \ttt{MSTP(4)}.
\begin{subentry}
\iteme{PARU(181) :} $\A^0$ coupling to down type quarks.
\iteme{PARU(182) :} $\A^0$ coupling to up type quarks.
\iteme{PARU(183) :} $\A^0$ coupling to leptons.
\iteme{PARU(184) :} $\A^0$ coupling to $\Z^0$.
\iteme{PARU(185) :} $\A^0$ coupling to $\W^{\pm}$.
\iteme{PARU(186) :} $\A^0$ coupling to $\Z^0 \hrm^0$ (or
$\Z^*$ to $\A^0 \hrm^0$), in MSSM $\cos(\beta-\alpha)$.
\iteme{PARU(187) :} $\A^0$ coupling to $\Z^0 \H^0$ (or $\Z^*$ to
$\A^0 \H^0$), in MSSM $\sin(\beta-\alpha)$.
\iteme{PARU(188) :} As \ttt{PARU(186)}, but coupling to $\Z'^0$
rather than $\Z^0$.
\iteme{PARU(189) :} As \ttt{PARU(187)}, but coupling to $\Z'^0$
rather than $\Z^0$.
\iteme{PARU(190) :} $\A^0$ coupling to $\H^{\pm}$ in
$\gamma \gamma \to \A^0$ loops, 0 in MSSM.
\end{subentry}
\iteme{PARU(191) - PARU(195) :} (D = 4*0., 1.) multiplicative factors
that can be used to modify the couplings of the $\H^{\pm}$ particle
in {\Py}. Currently only \ttt{PARU(195)} is in use. See above for
related comments.
\begin{subentry}
\iteme{PARU(195) :} $\H^{\pm}$ coupling to $\W^{\pm} \hrm^0$ (or
$\W^{* \pm}$ to $\H^{\pm} \hrm^0$), in MSSM $\cos(\beta-\alpha)$.
\end{subentry}
\iteme{PARU(197):} (D = 0.) $\H^0$ coupling to $\W^{\pm} \H^{\mp}$
within a two-Higgs-doublet model.
\iteme{PARU(198):} (D = 0.) $\A^0$ coupling to $\W^{\pm} \H^{\mp}$
within a two-Higgs-doublet model.
\boxsep
\iteme{PARJ(180) - PARJ(187) :}\label{p:PARJ180} couplings of the
second generation fermions to the $Z'^0$, following the same pattern
and with the same default values as the first one in
\ttt{PARU(121) - PARU(128)}.
\iteme{PARJ(188) - PARJ(195) :}couplings of the
third generation fermions to the $Z'^0$, following the same pattern
and with the same default values as the first one in
\ttt{PARU(121) - PARU(128)}.
\end{entry}
\drawbox{COMMON/PYTCSM/ITCM(0:99),RTCM(0:99)}\label{p:PYTCSM}
\begin{entry}
\itemc{Purpose:} to give access to a number of switches and parameters
which regulate the simulation of the TechniColor Strawman Model
\cite{Lan02,Lan02a}, plus a few further parameters related to
the simulation of compositeness, mainly in earlier incarnations of
TechniColor.
\boxsep
\iteme{ITCM(1) :}\label{p:ITCM} (D = 4) $N_{TC}$, number of technicolors;
fixes the relative values of $g_{\mrm{em}}$ and $g_{\mrm{etc}}$.
\iteme{ITCM(2) :} (D = 0) Topcolor model.
\begin{subentry}
\iteme{= 0 :} Standard Topcolor. Third generation quark couplings to
the coloron are proportional to $\cot\theta_3$, see \ttt{RTCM(21)}
below; first two generations are proportional to $-\tan\theta_3$.
\iteme{= 1 :} Flavor Universal Topcolor. All quarks couple with
strength proportional to $\cot\theta_3$.
\end{subentry}
\iteme{ITCM(5) :} (D = 0) presence of anomalous couplings in Standard Model
processes, see section \ref{sss:ancoupclass} for further details.
\begin{subentry}
\iteme{= 0 :} absent.
\iteme{= 1 :} left--left isoscalar model, with only $\u$ and $\d$
quarks composite (at the probed scale).
\iteme{= 2 :} left--left isoscalar model, with all quarks composite.
\iteme{= 3 :} helicity-non-conserving model, with only $\u$ and $\d$
quarks composite (at the probed scale).
\iteme{= 4 :} helicity-non-conserving model, with all quarks composite.
\iteme{= 5 :} coloured technihadrons, affecting the standard QCD
$2 \to 2$ cross sections by the exchange of Coloron or Colored Technirho,
see section \ref{sss:technicolorclass}.
\end{subentry}
\boxsep
\iteme{RTCM(1) :}\label{p:RTCM} (D = 82 GeV) $F_T$, the Technicolor decay
constant.
\iteme{RTCM(2) :} (D = 4/3) $Q_U$, charge of up-type technifermion;
the down-type technifermion has a charge $Q_D=Q_U-1$.
\iteme{RTCM(3) :} (D = 1/3) $\sin\chi$, where $\chi$ is the
mixing angle between isotriplet technipion interaction and mass
eigenstates.
\iteme{RTCM(4) :} (D = $1/\sqrt{6}$) $\sin\chi'$, where $\chi'$ is the
mixing angle between the isosinglet ${\pi'}^0_{\mrm{tc}}$ interaction
and mass eigenstates.
\iteme{RTCM(5) :} (D = 1) Clebsch for technipi decays to charm. Appears
squared in decay rates.
\iteme{RTCM(6) :} (D = 1) Clebsch for technipi decays to bottom. Appears
squared in decay rates.
\iteme{RTCM(7) :} (D = 0.0182) Clebsch for technipi decays to top,
estimated to be $m_{\b}/m_{\t}$. Appears squared in decay rates.
\iteme{RTCM(8) :} (D = 1) Clebsch for technipi decays to $\tau$.
Appears squared in decay rates.
\iteme{RTCM(9) :} (D = 0) squared Clebsch for isotriplet technipi decays
to gluons.
\iteme{RTCM(10) :} (D = 4/3) squared Clebsch for isosinglet technipi
decays to gluons.
\iteme{RTCM(11) :} (D = 0.05) technirho--techniomega mixing parameters.
Allows for isospin-violating decays of the techniomega.
\iteme{RTCM(12) :} (D = 200 GeV) vector technimeson decay parameter.
Affects the decay rates of vector technimesons into technipi plus
transverse gauge boson.
\iteme{RTCM(13) :} (D = 200 GeV) axial mass parameter for
technivector decays to transverse gauge bosons and technipions.
\iteme{RTCM(21) :} (D = $\sqrt{0.08}$) tangent of Topcolor mixing angle,
in the scenario with coloured technihadrons described in section
\ref{sss:technicolorclass} and switched on with \ttt{ITCM(5) = 5}.
For \ttt{ITCM(2) = 0}, the coupling of the $\mathrm{V}_8$ to light
quarks is suppressed by \ttt{RTCM(21)}$^2$ whereas the coupling to
heavy ($\b$ and $\t$) quarks is enhanced by 1/\ttt{RTCM(21)}$^2$.
For \ttt{ITCM(21) = 1}, the coupling to quarks
is universal, and given by 1/\ttt{RTCM(21)}$^2$.
\iteme{RTCM(22) :} (D = $1/\sqrt{2}$) sine of isosinglet technipi mixing
with Topcolor currents.
\iteme{RTCM(23) :} (D = 0) squared Clebsch for colour-octet technipi decays
to charm.
\iteme{RTCM(24) :} (D = 0) squared Clebsch for colour-octet technipi decays
to bottom.
\iteme{RTCM(25) :} (D = 0) squared Clebsch for colour-octet technipi decays
to top.
\iteme{RTCM(26) :} (D = 5/3) squared Clebsch for colour-octet technipi
decays to gluons.
\iteme{RTCM(27) :} (D = 250 GeV) colour-octet technirho decay parameter for
decays to technipi plus gluon.
\iteme{RTCM(28) :} (D = 250 GeV) hard mixing parameter between
colour-octet technirhos.
\iteme{RTCM(29) :} (D = $1/\sqrt{2}$) magnitude of $(1,1)$ element of the
\tbf{U(2)} matrices that diagonalize U-type technifermion condensates.
\iteme{RTCM(30) :} (D = 0 Radians) phase for the element described above,
\ttt{RTCM(29)}.
\iteme{RTCM(31) :} (D = $1/\sqrt{2}$) Magnitude of $(1,1)$ element of the
\tbf{U(2)} matrices that diagonalize D-type technifermion condensates.
\iteme{RTCM(32) :} (D = 0 Radians) phase for the element described above,
\ttt{RTCM(31)}.
\iteme{RTCM(33) :} (D = 1) if
$\Gamma_{V_8}(\hat{s}) > \mtt{RTCM(33)}\sqrt{\hat{s}}$, then
$\Gamma_{V_8}(\hat{s})$ is redefined to be $\mtt{RTCM(33)}\sqrt{\hat{s}}$.
It thus prevents the coloron from becoming wider than its mass.
\iteme{RTCM(41) :} (D = 1000 GeV) compositeness scale $\Lambda$, used in
processes involving excited fermions, and for Standard Model processes
when \ttt{ITCM(5)} is between 1 and 4.
\iteme{RTCM(42) :} (D = 1.) sign of the interference term between the
standard cross section and the compositeness term ($\eta$ parameter);
should be $\pm 1$; used for Standard Model processes when \ttt{ITCM(5)}
is between 1 and 4.
\iteme{RTCM(43) - RTCM(45) :} (D = 3*1.) strength of the \tbf{SU(2)},
\tbf{U(1)} and \tbf{SU(3)} couplings, respectively, in an excited fermion
scenario; cf. $f$, $f'$ and $f_s$ of \cite{Bau90}.
\iteme{RTCM(46) :} (D = 0.) anomalous magnetic moment of the $\W^{\pm}$
in process 20; $\eta = \kappa - 1$, where $\eta = 0$ ($\kappa = 1$)
is the Standard Model value.
\end{entry}
\subsection{Supersymmetry Common-Blocks and Routines}
\label{ss:susycode}
The parameters available to the SUSY user are stored in the
common block \ttt{PYMSSM}. In general, options are set by the \ttt{IMSS}
array, while real valued parameters are set by \ttt{RMSS}. The entries
\ttt{IMSS(0)} and \ttt{RMSS(0)} are not used, but are available
for compatibility with the C programming language. Note also that most
options are only used by {\Py}'s internal SUSY machinery and are ineffective
when external spectrum calculations are used, see section \ref{sss:models}.
\drawbox{COMMON/PYMSSM/IMSS(0:99),RMSS(0:99)}\label{p:PYMSSM}
\begin{entry}
\itemc{Purpose:} to give access to parameters that allow the
simulation of the MSSM.
\iteme{IMSS(1) :}\label{p:IMSS} (D = 0) level of MSSM simulation.
\begin{subentry}
\iteme{= 0 :} No MSSM simulation.
\iteme{= 1 :} A general MSSM simulation. The parameters of the model
are set by the array \ttt{RMSS}.
\iteme{= 2 :} An approximate SUGRA simulation using the analytic
formulae of \cite{Dre95} to reduce the number of free parameters.
In this case, only five input parameters are used.
\ttt{RMSS(1)} is the common gaugino mass
$m_{1/2}$, \ttt{RMSS(8)} is the common scalar mass $m_0$,
\ttt{RMSS(4)} fixes the sign of the higgsino mass $\mu$,
\ttt{RMSS(16)} is the common trilinear coupling $A$, and
\ttt{RMSS(5)} is $\tan\beta=v_2/v_1$.
\iteme{= 11 :} Read spectrum from a SUSY Les Houches Accord (SLHA)
conformant file. The Logical Unit Number on which the file is opened
should be put in \ttt{IMSS(21)}. If a decay table should also be read in,
the corresponding Unit Number (normally the same as the spectrum file)
should be put in \ttt{IMSS(22)}. Cross sections are
still calculated by {\Py}, as are decays for those sparticles and higgs
bosons for which a decay table is not found on the file.
\iteme{= 12 :} Invoke a runtime interface to \tsc{Isasusy} \cite{Bae93}
for determining SUSY mass spectrum and mixing parameters. This
provides a more precise solution of the renormalization group
equations than is offered by the option \ttt{= 2} above. The interface
automatically asks the \ttt{SUGRA} routine (part of \tsc{Isasusy}) to
solve the RGE's for the weak scale mass spectrum and mixing parameters.
The mSUGRA input parameters should be given in \ttt{RMSS} as usual, i.e.:
\ttt{RMSS(1) = }$m_{1/2}$, \ttt{RMSS(4) = }\sgnmu,
\ttt{RMSS(5) = }$\tan\beta$, \ttt{RMSS(8) = }$m_0$, and
\ttt{RMSS(16) = }$A$. As before, we are using the conventions of
\cite{Hab85, Gun86a} everywhere. Cross sections and decay widths are
still calculated by {\Py}, using the output provided by \tsc{Isasusy}.
Note that since {\Py} cannot always be expected to be linked with the
\tsc{Isajet} library, two dummy routines and a dummy function are
included. These are \ttt{SUGRA}, \ttt{SSMSSM} and \ttt{VISAJE}, located
towards the very bottom of the {\Py} source code. These routines must be
removed and {\Py} recompiled before a proper linking with \tsc{Isajet}
can be achieved. Furthermore, the common-block sizes and variable positions
accessed in the \ttt{SUGRA} routine have to match those of the \tsc{Isajet}
version used, see section \ref{sss:models}.
\iteme{= 13 :} File-based run-time \tsc{Isasusy} interface, i.e.\ using
an \tsc{Isajet} input file. The contents of the input file should be
identical to what would normally be typed when using the \tsc{Isajet}
RGE executable stand-alone (normally \ttt{isasugra.x}). The input file
should be opened by the user in his/her main program and the Logical
Unit Number should be stored in \ttt{IMSS(20)}, where {\Py} will look
for it during initialization. {\Py} will then pass the parameters to the
\ttt{SUGRA} subroutine in \tsc{Isajet} for RGE evolution and will
afterwards extract the electroweak scale mass and coupling spectra
to its own common blocks. For example, the first line of the input
file should contain the model code: 1 for mSUGRA, 2 for mGMSB, 3 for
non-universal SUGRA, 4 for SUGRA with truly unified gauge couplings,
5 for non-minimal GMSB, 6 for SUGRA+right-handed neutrino, 7 for
anomaly-mediated SUSY breaking. The ensuing lines should contain the
input parameters.\\
While option \ttt{IMSS(1) = 12} above can only be used for mSUGRA
scenarios, but then is easy to use, the current option allows the full
range of \tsc{Isajet} models to be accessed.
\end{subentry}
\iteme{IMSS(2) :} (D = 0) treatment of {\bf U(1)}, {\bf SU(2)}, and
{\bf SU(3)} gaugino mass parameters.
\begin{subentry}
\iteme{= 0 :} The gaugino parameters $M_1, M_2$ and $M_3$ are
set by \ttt{RMSS(1), RMSS(2),} and \ttt{RMSS(3)}, i.e.\ there is
no forced relation between them.
\iteme{= 1 :} The gaugino parameters are fixed by the relation
$(3/5) \, M_1/\alpha_1 =M_2/\alpha_2=M_3/\alpha_3=X$ and the
parameter \ttt{RMSS(1)}. If \ttt{IMSS(1) = 2}, then
\ttt{RMSS(1)} is treated as the common gaugino mass $m_{1/2}$
and \ttt{RMSS(20)} is the GUT scale coupling constant
$\alpha_{GUT}$, so that $X=m_{1/2}/\alpha_{GUT}$.
\iteme{= 2 :} $M_1$ is set by \ttt{RMSS(1)}, $M_2$ by \ttt{RMSS(2)}
and $M_3 = M_2\alpha_3/\alpha_2$. In such a scenario,
the {\bf U(1)} gaugino mass behaves anomalously.
\end{subentry}
\iteme{IMSS(3) :} (D = 0) treatment of the gluino mass parameter.
\begin{subentry}
\iteme{= 0 :} The gluino mass parameter $M_3$ is used to calculate
the gluino pole mass with the formulae of \cite{Kol96}. The effects of
squark loops can significantly shift the mass.
\iteme{= 1 :} $M_3$ is the gluino pole mass. The effects of squark
loops are assumed to have been included in this value.
\end{subentry}
\iteme{IMSS(4) :} (D = 1) treatment of the Higgs sector.
\begin{subentry}
\iteme{= 0 :} The Higgs sector is determined by
the approximate formulae of \cite{Car95} and the pseudoscalar mass
$M_{\A}$ set by \ttt{RMSS(19)}.
\iteme{= 1 :} The Higgs sector is determined by the exact
formulae of \cite{Car95} and the pseudoscalar mass $M_{\A}$ set by
\ttt{RMSS(19)}. The pole mass for $M_{\A}$ is not the same as the input
parameter.
\iteme{= 2 :} The Higgs sector is fixed by the mixing angle $\alpha$
set by \ttt{RMSS(18)} and the mass values \ttt{PMAS(I,1)}, where
\ttt{I = 25, 35, 36,} and \ttt{37}.
\iteme{= 3 :} Call \tsc{FeynHiggs} \cite{Hei99} for a precise calculation
of Higgs masses. For the time being, it can be invoked either when using
an SLHA SUSY spectrum, i.e.\ for \ttt{IMSS(1) = 11}, or when using the
run-time interface to \tsc{Isasusy}, i.e.\ for \ttt{IMSS(1) = 12, 13}.
When \tsc{FeynHiggs} is to be linked, the three dummy routines
\ttt{FHSETFLAGS}, \ttt{FHSETPARA} and \ttt{FHHIGGSCORR} need first be
removed from the {\Py} library.
\end{subentry}
\iteme{IMSS(5) :} (D = 0) allows you to set the $\st$, $\sbo$ and
$\stau$ masses and mixing by hand.
\begin{subentry}
\iteme{= 0 :} no, the program calculates itself.
\iteme{= 1 :} yes, calculate from given input. The parameters
\ttt{RMSS(26) - RMSS(28)} specify the mixing angle (in radians)
for the sbottom, stop, and stau. The parameters \ttt{RMSS(10) - RMSS(14)}
specify the two stop masses, the one sbottom mass (the other being fixed
by the other parameters) and the two stau masses. Note that the masses
\ttt{RMSS(10), RMSS(11)} and \ttt{RMSS(13)} correspond to the left-left
entries of the diagonalized matrices, while \ttt{RMSS(12)} and
\ttt{RMSS(14)} correspond to the right-right entries. Note that these
entries need not be ordered in mass.
\end{subentry}
\iteme{IMSS(7) :} (D = 0) treatment of the scalar masses in an
extension of SUGRA models. The presence of additional {\bf U(1)}
symmetries at high energy scales can modify the boundary
conditions for the scalar masses at the unification scale.
\begin{subentry}
\iteme{= 0 :} No additional $D$-terms are included. In
SUGRA models, all scalars have the mass $m_0$ at the
unification scale.
\iteme{= 1 :} \ttt{RMSS(23) - RMSS(25)} are the values of $D_X,
D_Y$ and $D_S$ at the unification scale in the model of \cite{Mar94}.
The boundary conditions for the scalar masses are shifted based
on their quantum numbers under the additional {\bf U(1)} symmetries.
\end{subentry}
\iteme{IMSS(8) :} (D = 0) treatment of the $\stau$ mass eigenstates.
\begin{subentry}
\iteme{= 0 :} The $\stau$ mass eigenstates are calculated
using the parameters\\
\ttt{RMSS(13, 14, 17)}.
\iteme{= 1 :} The $\stau$ mass eigenstates are identical to
the interaction eigenstates, so they are treated
identically to $\se$ and $\smu$ .
\end{subentry}
\iteme{IMSS(9) :} (D = 0) treatment of the right-handed squark mass
eigenstates for the first two generations.
\begin{subentry}
\iteme{= 0 :} The $\sq_R$ masses are fixed by \ttt{RMSS(9)}.
$\sd_R$ and $\su_R$ are identical except for Electroweak $D$-term
contributions.
\iteme{= 1 :} The masses of $\sd_R$ and $\su_R$
are fixed by \ttt{RMSS(9)} and \ttt{RMSS(22)} respectively.
\end{subentry}
\iteme{IMSS(10) :} (D = 0) allowed decays for $\chio_2$.
\begin{subentry}
\iteme{= 0 :} The second lightest neutralino $\chio_2$ decays
with a branching ratio calculated from the MSSM parameters.
\iteme{= 1 :} $\chio_2$ is forced to decay only to $\chio_1 \gamma$,
regardless of the actual branching ratio.
This can be used for detailed studies of this particular final state.
\end{subentry}
\iteme{IMSS(11) :} (D = 0) choice of the lightest superpartner (LSP).
\begin{subentry}
\iteme{= 0 :} $\chio_1$ is the LSP.
\iteme{= 1 :} $\chio_1$ is the next to lightest superparter (NLSP)
and the gravitino is the LSP. The $\chio_1$
decay length is calculated from the gravitino
mass set by \ttt{RMSS(21)} and the $\chio_1$ mass and mixing.
\end{subentry}
\iteme{IMSS(13) :} (D = 0) possibility to extend the particle content
recognized by {\Py} to that of the Next-to-Minimal Supersymmetric
Standard Model (NMSSM).
\begin{subentry}
\iteme{= 0 :} MSSM particle content.
\iteme{= 1 :} NMSSM particle content
\itemc{Note:} at present, \ttt{= 1} merely allows {\Py} to recognize
the NMSSM particles. {\Py} does not contain any internal machinery
for doing calculations in the NMSSM. Thus, the basic scattering
processes should be generated by an external program and handed to
{\Py} via the LHA interface for parton-level events.
This should then be combined with either setting the NMSSM resonance
decays by hand, or by reading in an SLHA decay table prepared by an
external decay package.
\end{subentry}
\iteme{IMSS(20) :} (D = 0) Logical Unit Number on which the SUSY model
parameter file is opened, for file-based run-time interface to the
\tsc{Isajet} SUSY RGE machinery, see \ttt{IMSS(1) = 13}.
\iteme{IMSS(21) :} (D = 0) Logical Unit Number for SLHA spectrum read-in.
Only used if \ttt{IMSS(1) = 11}.
\iteme{IMSS(22) :} (D = 0) Read-in of SLHA decay table.
\begin{subentry}
\iteme{= 0 :} No decays are read in. The internal {\Py} machinery is
used to calculate decay rates.
\iteme{> 0 :} Read decays from SLHA file on unit number
\ttt{IMSS(22)}. During initialization, decay tables in the file
will replace the values calculated by {\Py}. Particles for which
the file \emph{does not} contain a decay table will thus still
have their decays calculated by {\Py}. Any decay lines
associated with a zero width mother are ignored, giving a
fast way of switching off decays without having to comment out all
the decay lines.
In normal usage one would expect \ttt{IMSS(22)} to be equal to
\ttt{IMSS(21)}, to ensure that the spectrum and decays are consistent
with each other, but this is not a strict requirement.
\end{subentry}
\iteme{IMSS(23) :} (D = 0) writing of MSSM spectrum data.
\begin{subentry}
\iteme{= 0 :} Don't write out spectrum.
\iteme{> 0 :} Write out spectrum in SLHA format (calculated by {\Py} or
otherwise) to file on unit number \ttt{IMSS(23)}.
\end{subentry}
\iteme{IMSS(24) :} (D = 0) writing of MSSM particle decay table.
\begin{subentry}
\iteme{= 0 :} Don't write out decay table.
\iteme{> 0 :} Write out decay table in SLHA format to file on unit number
\ttt{IMSS(24)}. Not implemented in the code yet.
In normal usage one would expect \ttt{IMSS(24)} to be equal to
\ttt{IMSS(23)}, to ensure that the spectrum and decays are consistent
with each other, but this is not a strict requirement.
\end{subentry}
\iteme{IMSS(51) :} (D = 0) Lepton number violation on/off
(LLE type couplings).
\begin{subentry}
\iteme{= 0 :} All LLE couplings off. LLE decay channels off.
\iteme{= 1 :} All LLE couplings set to common value given by
$10^\ttt{\scriptsize-RMSS(51)}$.
\iteme{= 2 :} LLE couplings set to generation-hierarchical
`natural' values with common normalization \ttt{RMSS(51)}
(see section \ref{sss:rparityviol}).
\iteme{= 3 :} All LLE couplings set to zero, but LLE decay channels not
switched off. Non-zero couplings should be entered individually into the
array \ttt{RVLAM(I,J,K)}. Because of the antisymmetry in I and J, only
entries with I $<$ J need be entered.
\end{subentry}
\iteme{IMSS(52) :} (D = 0) Lepton number violation on/off
(LQD type couplings).
\begin{subentry}
\iteme{= 0 :} All LQD couplings off. LQD decay channels off.
\iteme{= 1 :} All LQD couplings set to common value given by
$10^\ttt{\scriptsize-RMSS(52)}$.
\iteme{= 2 :} LQD couplings set to generation-hierarchical
`natural' values with common normalization \ttt{RMSS(52)}
(see section \ref{sss:rparityviol}).
\iteme{= 3 :} All LQD couplings set to zero, but LQD decay channels not
switched off. Non-zero couplings should be entered individually into the
array \ttt{RVLAMP(I,J,K)}.
\end{subentry}
\iteme{IMSS(53) :} (D = 0) Baryon number violation on/off
\begin{subentry}
\iteme{= 0 :} All UDD couplings off. UDD decay channels off.
\iteme{= 1 :} All UDD couplings set to common value given by
$10^\ttt{\scriptsize-RMSS(53)}$.
\iteme{= 2 :} UDD couplings set to generation-hierarchical
`natural' values with common normalization \ttt{RMSS(53)}
(see section \ref{sss:rparityviol}).
\iteme{= 3 :} All UDD couplings set to zero, but UDD decay channels not
switched off. Non-zero couplings should be entered individually into the
array \ttt{RVLAMB(I,J,K)}. Because of the antisymmetry in J and K, only
entries with J $<$ K need be entered.
\end{subentry}
\boxsep
\iteme{RMSS(1) :}\label{p:RMSS} (D = 80. GeV) If \ttt{IMSS(1) = 1} $M_1$,
then {\bf U(1)} gaugino mass. If \ttt{IMSS(1) = 2}, then the common
gaugino mass $m_{1/2}$.
\iteme{RMSS(2) :} (D = 160. GeV) $M_2$, the {\bf SU(2)} gaugino mass.
\iteme{RMSS(3) :} (D = 500. GeV) $M_3$, the {\bf SU(3)} (gluino)
mass parameter.
\iteme{RMSS(4) :} (D = 800. GeV) $\mu$, the higgsino mass parameter.
If \ttt{IMSS(1) = 2}, only the sign of $\mu$ is used.
\iteme{RMSS(5) :} (D = 2.) $\tan\beta$, the ratio of Higgs expectation
values.
\iteme{RMSS(6) :} (D = 250. GeV) Left slepton mass $M_{\sell_L}$.
The sneutrino mass is fixed by a sum rule.
\iteme{RMSS(7) :} (D = 200. GeV) Right slepton mass $M_{\sell_R}$.
\iteme{RMSS(8) :} (D = 800. GeV) Left squark mass $M_{\sq_L}$. If
\ttt{IMSS(1) = 2}, the common scalar mass $m_0$.
\iteme{RMSS(9) :} (D = 700. GeV) Right squark mass $M_{\sq_R}$. $M_{\sd_R}$
when \ttt{IMSS(9) = 1}.
\iteme{RMSS(10) :} (D = 800. GeV) Left squark mass for the third
generation $M_{\sq_L}$. When \ttt{IMSS(5) = 1}, it is instead the
$\st_2$ mass, and $M_{\sq_L}$ is a derived quantity.
\iteme{RMSS(11) :} (D = 700. GeV) Right sbottom mass $M_{\sbo_R}$. When
\ttt{IMSS(5) = 1}, it is instead the $\sbo_1$ mass.
\iteme{RMSS(12) :} (D = 500. GeV) Right stop mass $M_{\st_R}$ If
negative, then it is assumed that $M_{\st_R}^2 < 0$. When
\ttt{IMSS(5) = 1 }, it is instead the $\st_1$ mass.
\iteme{RMSS(13) :} (D = 250. GeV) Left stau mass $M_{\stau_L}$.
\iteme{RMSS(14) :} (D = 200. GeV) Right stau mass $M_{\stau_R}$.
\iteme{RMSS(15) :} (D = 800. GeV) Bottom trilinear coupling $A_{\b}$. When
\ttt{IMSS(5) = 1}, it is a derived quantity.
\iteme{RMSS(16) :} (D = 400. GeV) Top trilinear coupling $A_{\t}$. If
\ttt{IMSS(1) = 2}, the common trilinear coupling $A$. When
\ttt{IMSS(5) = 1}, it is a derived quantity.
\iteme{RMSS(17) :} (D = 0.) Tau trilinear coupling $A_{\tau}$. When
\ttt{IMSS(5) = 1}, it is a derived quantity.
\iteme{RMSS(18) :} (D = 0.1) Higgs mixing angle $\alpha$. This is only
used when all of the Higgs parameters are set by you, i.e
\ttt{IMSS(4) = 2}.
\iteme{RMSS(19) :} (D = 850. GeV) Pseudoscalar Higgs mass parameter $M_{\A}$.
\iteme{RMSS(20) :} (D = 0.041) GUT scale coupling constant
$\alpha_{\mrm{GUT}}$.
\iteme{RMSS(21) :} (D = 1.0 eV) The gravitino mass. Note nonconventional
choice of units for this particular mass.
\iteme{RMSS(22) :} (D = 800. GeV) $\su_R$ mass when \ttt{IMSS(9) = 1}.
\iteme{RMSS(23) :} (D = 10$^4$ GeV$^2$) $D_X$ contribution to scalar
masses when \ttt{IMSS(7) = 1}.
\iteme{RMSS(24) :} (D = 10$^4$ GeV$^2$) $D_Y$ contribution to scalar
masses when \ttt{IMSS(7) = 1}.
\iteme{RMSS(25) :} (D = 10$^4$ GeV$^2$) $D_S$ contribution to scalar
masses when \ttt{IMSS(7) = 1}.
\iteme{RMSS(26) :} (D = 0.0 radians) when \ttt{IMSS(5) = 1} it is the
sbottom mixing angle.
\iteme{RMSS(27) :} (D = 0.0 radians) when \ttt{IMSS(5) = 1} it is the
stop mixing angle.
\iteme{RMSS(28) :} (D = 0.0 radians) when \ttt{IMSS(5) = 1} it is the
stau mixing angle.
\iteme{RMSS(29) :} (D = $2.4 \times 10^{18}$ GeV) The Planck mass,
used for calculating decays to light gravitinos.
\iteme{RMSS(30) - RMSS(33) :} (D = 0.0, 0.0, 0.0, 0.0) complex phases for
the mass parameters in \ttt{RMSS(1) - RMSS(4)}, where the latter represent
the moduli of the mass parameters for the case of nonvanishing phases.
\iteme{RMSS(40), RMSS(41) :} used for temporary storage of the corrections
$\Delta m_{\t}$ and $\Delta m_{\b}$, respectively, in the calculation of
Higgs properties.
\iteme{RMSS(51) :} (D = 0.0) when \ttt{IMSS(51) = 1} it is the negative
logarithm of the common value for all lepton-number-violating
$\lambda$ couplings (LLE). When \ttt{IMSS(51) = 2} it is the constant of
proportionality for generation-hierarchical $\lambda$ couplings. See
section \ref{sss:rparityviol}.
\iteme{RMSS(52) :} (D = 0.0) when \ttt{IMSS(52) = 1} it is the negative
logarithm of the common value for all lepton-number-violating
$\lambda'$ couplings (LQD). When \ttt{IMSS(52) = 2} it is the constant of
proportionality for generation-hierarchical $\lambda'$ couplings. See
section \ref{sss:rparityviol}.
\iteme{RMSS(53) :} (D = 0.0) when \ttt{IMSS(53) = 1} it is the negative
logarithm of the common value for all baryon-number-violating
$\lambda''$ couplings (UDD). When \ttt{IMSS(53) = 2} it is the constant of
proportionality for generation-hierarchical $\lambda''$ couplings. See
section \ref{sss:rparityviol}.
\end{entry}
\drawboxtwo{~COMMON/PYSSMT/ZMIX(4,4),UMIX(2,2),VMIX(2,2),SMZ(4),SMW(2),}%
{\&SFMIX(16,4),ZMIXI(4,4),UMIXI(2,2),VMIXI(2,2)}\label{p:PYSSMT}
\begin{entry}
\itemc{Purpose:} to provide information on the neutralino, chargino,
and sfermion mixing parameters. The variables should not be changed
by you.
\iteme{ZMIX(4,4) :}\label{p:ZMIX} the real part of the neutralino mixing
matrix in the Bino--neutral Wino--Up higgsino--Down higgsino basis.
\iteme{UMIX(2,2) :}\label{p:UMIX} the real part of the chargino mixing
matrix in the charged Wino--charged higgsino basis.
\iteme{VMIX(2,2) :}\label{p:VMIX} the real part of the charged conjugate
chargino mixing matrix in the wino--charged higgsino basis.
\iteme{SMZ(4) :}\label{p:SMZ} the signed masses of the neutralinos.
\iteme{SMW(2) :}\label{p:SMW} the signed masses of the charginos.
\iteme{SFMIX(16,4) :}\label{p:SFMIX} the sfermion mixing matrices
\tbf{T} in the L--R basis, identified by the corresponding fermion, i.e.\
\ttt{SFMIX(6,I)} is the stop mixing matrix. The four entries for each
sfermion are $\mrm{T}_{11}, \mrm{T}_{12}, \mrm{T}_{21},$ and
$\mrm{T}_{22}$.
\iteme{ZMIXI(4,4) :}\label{p:ZMIXI} the imaginary part of the neutralino
mixing matrix in the Bino--neutral Wino--Up higgsino--Down higgsino basis.
\iteme{UMIXI(2,2) :}\label{p:UMIXI} the imaginary part of the chargino
mixing matrix in the charged Wino--charged higgsino basis.
\iteme{VMIXI(2,2) :}\label{p:VMIXI} the imaginary part of the charged
conjugate chargino mixing matrix in the wino--charged higgsino basis.
\end{entry}
\drawbox{~COMMON/PYMSRV/RVLAM(3,3,3), RVLAMP(3,3,3), RVLAMB(3,3,3)}%
\label{p:PYMSRV}
\begin{entry}
\itemc{Purpose:} to provide information on lepton- and
baryon-number-violating couplings.
\iteme{RVLAM(3,3,3) :}\label{p:RVLAM} the lepton-number-violating
$\lambda_{ijk}$ couplings. See \ttt{IMSS(51)}, \ttt{RMSS(51)}.
\iteme{RVLAMP(3,3,3) :}\label{p:RVLAMP} the lepton-number-violating
$\lambda'_{ijk}$ couplings. See \ttt{IMSS(52)}, \ttt{RMSS(52)}.
\iteme{RVLAMB(3,3,3) :}\label{p:RVLAMB} the baryon-number-violating
$\lambda''_{ijk}$ couplings. See \ttt{IMSS(53)}, \ttt{RMSS(53)}.
\end{entry}
\boxsep
The following subroutines and functions need not be accessed by the
user, but are described for completeness.
\begin{entry}
\iteme{SUBROUTINE PYAPPS :} uses approximate analytic formulae to
determine the full set of MSSM parameters from SUGRA inputs.
\iteme{SUBROUTINE PYGLUI :} calculates gluino decay modes.
\iteme{SUBROUTINE PYGQQB :} calculates three-body decays of gluinos
into neutralinos or charg\-inos and third generation fermions. These
routines are valid for large values of $\tan\beta$.
\iteme{SUBROUTINE PYCJDC :} calculates the chargino decay modes.
\iteme{SUBROUTINE PYHEXT :} calculates the non--Standard Model decay
modes of the Higgs bosons.
\iteme{SUBROUTINE PYHGGM :} determines the Higgs boson mass spectrum
using several inputs.
\iteme{SUBROUTINE PYINOM :} finds the mass eigenstates and mixing
matrices for the charginos and neutralinos.
\iteme{SUBROUTINE PYMSIN :} initializes the MSSM simulation.
\iteme{SUBROUTINE PYSLHA :} to read in or write out SUSY Les Houches
Accord spectra and decay tables. Can also be used stand-alone, before
the call to \ttt{PYINIT}, to read in SLHA decay tables for specific
particles. See section \ref{ss:parapartdat} for how to do this.
\iteme{SUBROUTINE PYNJDC :} calculates neutralino decay modes.
\iteme{SUBROUTINE PYPOLE :} computes the Higgs boson masses using
a renormalization group improved leading-log approximation and
two-loop leading-log corrections.
\iteme{SUBROUTINE PYSFDC :} calculates sfermion decay modes.
\iteme{SUBROUTINE PYSUBH :} computes the Higgs boson masses using
only renormalization group improved formulae.
\iteme{SUBROUTINE PYTBDY :} samples the phase space for three-body
decays of neutralinos, charginos, and the gluino.
\iteme{SUBROUTINE PYTHRG :} computes the masses and mixing matrices of
the third generation sfermions.
\iteme{SUBROUTINE PYRVSF :} $R$-violating sfermion decay widths.
\iteme{SUBROUTINE PYRVNE :} $R$-violating neutralino decay widths.
\iteme{SUBROUTINE PYRVCH :} $R$-violating chargino decay widths.
\iteme{SUBROUTINE PYRVGW :} calculates $R$-violating 3-body widths using
\ttt{PYRVI1, PYRVI2, PYRVI3, PYRVG1, PYRVG2, PYRVG3, PYRVG4, PYRVR}, and
\ttt{PYRVS}.
\iteme{FUNCTION PYRVSB :} calculates $R$-violating 2-body widths.
\iteme{SUBROUTINE SUGRA :} dummy routine, to avoid linking problems when
\tsc{Isajet} is not linked; see \ttt{IMSS(1) = 12}.
\iteme{SUBROUTINE SSMSSM :} dummy routine, to avoid linking problems when
\tsc{Isajet} is not linked; see \ttt{IMSS(1) = 12}.
\iteme{FUNCTION VISAJE :} dummy routine, to avoid linking problems when
\tsc{Isajet} is not linked; see \ttt{IMSS(1) = 12}.
\end{entry}
\subsection{General Event Information}
When an event is generated with \ttt{PYEVNT}, some information on
it is stored in the \ttt{MSTI} and \ttt{PARI} arrays of the
\ttt{PYPARS} common block (often copied directly from the internal
\ttt{MINT} and \ttt{VINT} variables). Further information is stored
in the complete event record; see section \ref{ss:evrec}.
Part of the information is only relevant for some subprocesses; by
default everything irrelevant is set to 0. Kindly note that, like the
\ttt{CKIN} constraints described in section \ref{ss:PYswitchkin},
kinematical variables normally (i.e.\ where it is not explicitly stated
otherwise) refer to the na\"{\i}ve hard scattering, before initial- and
final-state radiation effects have been included.
\drawbox{COMMON/PYPARS/MSTP(200),PARP(200),MSTI(200),PARI(200)}%
\label{p:PYPARS2}
\begin{entry}
\itemc{Purpose:} to provide information on latest event generated or,
in a few cases, on statistics accumulated during the run.
\iteme{MSTI(1) :}\label{p:MSTI} specifies the general type of
subprocess that has occurred, according to the {\ISUB} code given
in section \ref{ss:ISUBcode}.
\iteme{MSTI(2) :} whenever \ttt{MSTI(1)} (together with \ttt{MSTI(15)}
and \ttt{MSTI(16)}) are not enough to specify the type of process
uniquely, \ttt{MSTI(2)} provides an ordering of the different
possibilities. This is particularly relevant for the different
colour-flow topologies possible in QCD $2 \to 2$ processes, but
easily generalizes e.g.\ if a quark is replaced by a squark. With
$i = $\ttt{MSTI(15)}, $j = $\ttt{MSTI(16)} and $k = $\ttt{MSTI(2)},
the QCD possibilities are, in the classification scheme of
\cite{Ben84} (cf. section \ref{sss:QCDjetclass}):
\begin{subentry}
\itemn{ISUB = 11,} $i = j$, $\q_i \q_i \to \q_i \q_i$; \\
$k = 1$ : colour configuration $A$. \\
$k = 2$ : colour configuration $B$.
\itemn{ISUB = 11,} $i \neq j$, $\q_i \q_j \to \q_i \q_j$; \\
$k = 1$ : only possibility.
\itemn{ISUB = 12,} $\q_i \qbar_i \to \q_l \qbar_l$; \\
$k = 1$ : only possibility.
\itemn{ISUB = 13,} $\q_i \qbar_i \to \g \g$; \\
$k = 1$ : colour configuration $A$. \\
$k = 2$ : colour configuration $B$.
\itemn{ISUB = 28,} $\q_i \g \to \q_i \g$; \\
$k = 1$ : colour configuration $A$. \\
$k = 2$ : colour configuration $B$.
\itemn{ISUB = 53,} $\g \g \to \q_l \qbar_l$; \\
$k = 1$ : colour configuration $A$. \\
$k = 2$ : colour configuration $B$.
\itemn{ISUB = 68,} $\g \g \to \g \g$; \\
$k = 1$ : colour configuration $A$. \\
$k = 2$ : colour configuration $B$. \\
$k = 3$ : colour configuration $C$.
\itemn{ISUB = 83,} $\f \q \to \f' \Q$ (by $t$-channel $\W$ exchange;
does not distinguish colour flows but result of user selection); \\
$k = 1$ : heavy flavour $\Q$ is produced on side 1. \\
$k = 2$ : heavy flavour $\Q$ is produced on side 2.
\end{subentry}
\iteme{MSTI(3) :} the number of partons produced in the hard
interactions, i.e.\ the number $n$ of the $2 \to n$ matrix elements
used; it is sometimes 3 or 4 when a basic $2 \to 1$ or $2 \to 2$
process has been folded with two $1 \to 2$ initial branchings (like
$\q_i \q_j \to \q_k \q_l \hrm^0$).
\iteme{MSTI(4) :} number of documentation lines at the beginning of
the common block \ttt{PYJETS} that are given with \ttt{K(I,1) = 21};
0 for \ttt{MSTP(125) = 0}.
\iteme{MSTI(5) :} number of events generated to date in current
run. In runs with the variable-energy option, \ttt{MSTP(171) = 1}
and \ttt{MSTP(172) = 2}, only those events that survive (i.e.\ that
do not have \ttt{MSTI(61) = 1}) are counted in this number. That
is, \ttt{MSTI(5)} may be less than the total number of \ttt{PYEVNT}
calls.
\iteme{MSTI(6) :} current frame of event, cf. \ttt{MSTP(124)}.
\iteme{MSTI(7), MSTI(8) :} line number for documentation of
outgoing partons/particles from hard scattering for $2 \to 2$ or
$2 \to 1 \to 2$ processes (else = 0).
\iteme{MSTI(9) :} event class used in current event for $\gamma\p$ or
$\gamma\gamma$ events. The code depends on which process is being studied.
\begin{subentry}
\iteme{= 0 :} for other processes than the ones listed below.
\itemn{For $\gamma\p$ or $\gast\p$ events,} generated with the
\ttt{MSTP(14) = 10} or \ttt{MSTP(14) = 30} options:
\iteme{= 1 :} VMD.
\iteme{= 2 :} direct.
\iteme{= 3 :} anomalous.
\iteme{= 4 :} DIS (only for $\gamma^*\p$, i.e. \ttt{MSTP(14) = 30}).
\itemn{For real $\gamma\gamma$ events,} i.e.\ \ttt{MSTP(14) = 10}:
\iteme{= 1 :} VMD$\times$VMD.
\iteme{= 2 :} VMD$\times$direct.
\iteme{= 3 :} VMD$\times$anomalous .
\iteme{= 4 :} direct$\times$direct.
\iteme{= 5 :} direct$\times$anomalous.
\iteme{= 6 :} anomalous$\times$anomalous.
\itemn{For virtual $\gast\gast$ events,} i.e.\ \ttt{MSTP(14) = 30},
where the two incoming photons are not equivalent and the order
therefore matters:
\iteme{= 1 :} direct$\times$direct.
\iteme{= 2 :} direct$\times$VMD.
\iteme{= 3 :} direct$\times$anomalous.
\iteme{= 4 :} VMD$\times$direct.
\iteme{= 5 :} VMD$\times$VMD.
\iteme{= 6 :} VMD$\times$anomalous.
\iteme{= 7 :} anomalous$\times$direct.
\iteme{= 8 :} anomalous$\times$VMD.
\iteme{= 9 :} anomalous$\times$anomalous.
\iteme{= 10 :} DIS$\times$VMD.
\iteme{= 11 :} DIS$\times$anomalous.
\iteme{= 12 :} VMD$\times$DIS.
\iteme{= 13 :} anomalous$\times$DIS.
\end{subentry}
\iteme{MSTI(10) :} is 1 if cross section maximum was violated
in current event, and 0 if not.
\iteme{MSTI(11) :} {\KF} flavour code for beam (side 1) particle.
\iteme{MSTI(12) :} {\KF} flavour code for target (side 2) particle.
\iteme{MSTI(13), MSTI(14) :} {\KF} flavour codes for side 1 and side 2
initial-state shower initiators.
\iteme{MSTI(15), MSTI(16) :} {\KF} flavour codes for side 1 and side 2
incoming partons to the hard interaction.
\iteme{MSTI(17), MSTI(18) :} flag to signal if particle on side
1 or side 2 has been scattered diffractively; 0 if no, 1 if yes.
\iteme{MSTI(21) - MSTI(24) :} {\KF} flavour codes for outgoing partons
from the hard interaction. The number of positions actually used is
process-dependent, see \ttt{MSTI(3)}; trailing positions not used
are set = 0. For events with many outgoing partons, e.g.\ in external
processes, also \ttt{MSTI(25)} and \ttt{MSTI(26)} could be used.
\iteme{MSTI(25), MSTI(26) :} {\KF} flavour codes of the products in the
decay of a single $s$-channel resonance formed in the hard interaction.
Are thus only used when \ttt{MSTI(3) = 1} and the resonance is allowed
to decay.
\iteme{MSTI(31) :} number of hard or semi-hard scatterings that occurred
in the current event in the multiple-interaction scenario; is = 0 for a
low-$\pT$ event.
\iteme{MSTI(32) :} information on whether a reconnection occurred in the
current event; is 0 normally but 1 in case of reconnection.
\iteme{MSTI(41) :} the number of pile-up events generated in the latest
\ttt{PYEVNT} call (including the first, `hard' event).
\iteme{MSTI(42) - MSTI(50) :} {\ISUB} codes for the events 2--10
generated in the pile-up-events scenario. The first event {\ISUB} code is
stored in \ttt{MSTI(1)}. If \ttt{MSTI(41)} is less than 10, only as
many positions are filled as there are pile-up events. If MSTI(41) is
above 10, some {\ISUB} codes will not appear anywhere.
\iteme{MSTI(51) :} normally 0 but set to 1 if a \ttt{UPEVNT} call
did not return an event, such that \ttt{PYEVNT} could not generate
an event. For further details, see section \ref{ss:PYnewproc}.
\iteme{MSTI(52) :} counter for the number of times the current event
configuration failed in the generation machinery. For accepted events
this is always 0, but the counter can be used inside \ttt{UPEVNT}
to check on anomalous occurrences. For further
details, see section \ref{ss:PYnewproc}.
\iteme{MSTI(53) :} normally 0, but 1 if no processes with non-vanishing
cross sections were found in a \ttt{PYINIT} call, for the case that
\ttt{MSTP(127) = 1}.
\iteme{MSTI(61) :} status flag set when events are generated. It is
only of interest for runs with variable energies, \ttt{MSTP(171) = 1},
with the option \ttt{MSTP(172) = 2}.
\begin{subentry}
\iteme{= 0 :} an event has been generated.
\iteme{= 1 :} no event was generated, either because the c.m.\ energy
was too low or because the Monte Carlo phase space point selection
machinery rejected the trial point. A new energy is to be picked by
you.
\end{subentry}
\iteme{MSTI(71), MSTI(72) :} {\KF} code for incoming lepton beam or target
particles, when a flux of virtual photons are generated internally for
{\galep} beams, while \ttt{MSTI(11)} and \ttt{MSTI(12)} is
then the photon code.
\boxsep
\iteme{PARI(1) :}\label{p:PARI} total integrated cross section for
the processes
under study, in mb. This number is obtained as a by-product of the
selection of hard-process kinematics, and is thus known with better
accuracy when more events have been generated. The value stored here
is based on all events until the latest one generated.
\iteme{PARI(2) :} for unweighted events, \ttt{MSTP(142) = 0} or \ttt{= 2},
it is the ratio \ttt{PARI(1)/MSTI(5)}, i.e.\ the ratio of total
integrated cross section and number of events generated. Histograms
should then be filled with unit event weight and, at the end of the run,
multiplied by \ttt{PARI(2)} and divided by the bin width to convert
results to mb/(dimension of the horizontal axis).
For weighted events, \ttt{MSTP(142) = 1}, \ttt{MSTI(5)} is replaced by the
sum of \ttt{PARI(10)} values. Histograms should then be filled with
event weight \ttt{PARI(10)} and, as before, be multiplied by \ttt{PARI(2)}
and divided by the bin width at the end of the run.
In runs with the variable-energy option, \ttt{MSTP(171) = 1}
and \ttt{MSTP(172) = 2}, only those events that survive (i.e.\ that
do not have \ttt{MSTI(61) = 1}) are counted.
\iteme{PARI(7) :} an event weight, normally 1 and thus uninteresting,
but for external processes with \ttt{IDWTUP = -1, -2} or \ttt{-3} it can
be $-1$ for events with negative cross section, with \ttt{IDWTUP = 4} it
can be an arbitrary non-negative weight of dimension mb, and with
\ttt{IDWTUP = -4} it can be an arbitrary weight of dimension mb.
(The difference being that in most cases a rejection step is involved
to bring the accepted events to a common weight normalization, up to
a sign, while no rejection need be involved in the last two cases.)
\iteme{PARI(9) :} is weight \ttt{WTXS} returned from \ttt{PYEVWT}
call when \ttt{MSTP(142)} $\geq 1$, otherwise is 1.
\iteme{PARI(10) :} is compensating weight \ttt{1./WTXS} that should
be associated to events when \ttt{MSTP(142) = 1}, else is 1.
\iteme{PARI(11) :} $E_{\mrm{cm}}$, i.e.\ total c.m.\ energy
(except when using the {\galep} machinery, see \ttt{PARI(101)}.
\iteme{PARI(12) :} $s$, i.e.\ squared total c.m.\ energy
(except when using the {\galep} machinery, see \ttt{PARI(102)}.
\iteme{PARI(13) :} $\hat{m} = \sqrt{\hat{s}}$, i.e.\ mass of the
hard-scattering subsystem.
\iteme{PARI(14) :} $\hat{s}$ of the hard subprocess
($2 \to 2$ or $2 \to 1$).
\iteme{PARI(15) :} $\hat{t}$ of the hard subprocess
($2 \to 2$ or $2 \to 1 \to 2$).
\iteme{PARI(16) :} $\hat{u}$ of the hard subprocess
($2 \to 2$ or $2 \to 1 \to 2$).
\iteme{PARI(17) :} $\hat{p}_{\perp}$ of the hard subprocess
($2 \to 2$ or $2 \to 1 \to 2$),
evaluated in the rest frame of the hard interaction.
\iteme{PARI(18) :} $\hat{p}_{\perp}^2$ of the hard subprocess;
see \ttt{PARI(17)}.
\iteme{PARI(19) :} $\hat{m}'$, the mass of the complete three- or
four-body final state in $2 \to 3$ or $2 \to 4$ processes (while
$\hat{m}$, given in \ttt{PARI(13)}, here corresponds to the one-
or two-body central system).
Kinematically $\hat{m} \leq \hat{m}' \leq E_{\mrm{cm}}$.
\iteme{PARI(20) :} $\hat{s}' = \hat{m}'^2$; see \ttt{PARI(19)}.
\iteme{PARI(21) :} $Q$ of the hard-scattering subprocess. The exact
definition is process-dependent, see \ttt{MSTP(32)}.
\iteme{PARI(22) :} $Q^2$ of the hard-scattering subprocess; see
\ttt{PARI(21)}.
\iteme{PARI(23) :} $Q$ of the outer hard-scattering subprocess.
Agrees with \ttt{PARI(21)} for a $2 \to 1$ or $2 \to 2$ process.
For a $2 \to 3$ or $2 \to 4$ $\W/\Z$ fusion process, it is set by
the $\W/\Z$ mass scale, and for subprocesses 121 and 122 by the
heavy-quark mass.
\iteme{PARI(24) :} $Q^2$ of the outer hard-scattering subprocess;
see \ttt{PARI(23)}.
\iteme{PARI(25) :} $Q$ scale used as maximum virtuality in parton
showers. Is equal to \ttt{PARI(23)}, except for
Deeply Inelastic Scattering processes when \ttt{MSTP(22)} $\geq 1$.
\iteme{PARI(26) :} $Q^2$ scale in parton showers; see \ttt{PARI(25)}.
\iteme{PARI(31), PARI(32) :} the momentum fractions $x$ of the
initial-state parton-shower initiators on side 1 and 2, respectively.
\iteme{PARI(33), PARI(34) :} the momentum fractions $x$ taken by the
partons at the hard interaction, as used e.g.\ in the
parton-distribution functions.
\iteme{PARI(35) :} Feynman-$x$,
$x_{\mrm{F}} = x_1 - x_2 = $\,\ttt{PARI(33)}$-$\ttt{PARI(34)}.
\iteme{PARI(36) :}
$\tau = \hat{s}/s = x_1 \, x_2 = $\,\ttt{PARI(33)}$\times$\ttt{PARI(34)}.
\iteme{PARI(37) :} $y = (1/2) \ln(x_1/x_2)$, i.e.\ rapidity of the
hard-interaction subsystem in the c.m.\ frame of the event as a whole.
\iteme{PARI(38) :} $\tau' = \hat{s}'/s = $\,\ttt{PARI(20)}$/$\ttt{PARI(12)}.
\iteme{PARI(39), PARI(40) :} the primordial $k_{\perp}$ values
selected in the two beam remnants.
\iteme{PARI(41) :} $\cos\hat{\theta}$, where $\hat{\theta}$ is the
scattering angle of a $2 \to 2$ (or $2 \to 1 \to 2$) interaction,
defined in the rest frame of the hard-scattering subsystem.
\iteme{PARI(42) :} $x_{\perp}$, i.e.\ scaled transverse momentum of the
hard-scattering subprocess,
$x_{\perp} = 2 \hat{p}_{\perp}/E_{\mrm{cm}} = 2$\,%
\ttt{PARI(17)}$/$\ttt{PARI(11)}.
\iteme{PARI(43), PARI(44) :} $x_{L3}$ and $x_{L4}$, i.e.\ longitudinal
momentum fractions of the two scattered partons, in the range
$-1 < x_{\mrm{L}} < 1$, in the c.m.\ frame of the event as a whole.
\iteme{PARI(45), PARI(46) :} $x_3$ and $x_4$, i.e.\ scaled energy
fractions of the two scattered partons, in the c.m.\ frame of the
event as a whole.
\iteme{PARI(47), PARI(48) :} $y^*_3$ and $y^*_4$, i.e.\ rapidities
of the two scattered partons in the c.m.\ frame of the
event as a whole.
\iteme{PARI(49), PARI(50) :} $\eta^*_3$ and $\eta^*_4$, i.e.\
pseudorapidities of the two scattered partons in the c.m.\ frame
of the event as a whole.
\iteme{PARI(51), PARI(52) :} $\cos\theta^*_3$ and $\cos\theta^*_4$,
i.e.\ cosines of the polar angles of the two scattered partons in
the c.m.\ frame of the event as a whole.
\iteme{PARI(53), PARI(54) :} $\theta^*_3$ and $\theta^*_4$, i.e.\
polar angles of the two scattered partons, defined in the range
$0 < \theta^* < \pi$, in the c.m.\ frame of the event as a whole.
\iteme{PARI(55), PARI(56) :} azimuthal angles $\phi^*_3$ and
$\phi^*_4$ of the two scattered partons, defined in the range
$-\pi < \phi^* < \pi$, in the c.m.\ frame of the event as a whole.
\iteme{PARI(61) :} multiple interaction enhancement factor for
current event. A large value corresponds to a central collision
and a small value to a peripheral one.
\iteme{PARI(65) :} sum of the transverse momenta of partons
generated at the hardest interaction of the event, excluding
initial- and final-state radiation, i.e.\ $2 \times$\ttt{PARI(17)}.
Only intended for $2 \to 2$ or $2 \to 1 \to 2$ processes,
i.e.\ not implemented for $2 \to 3$ ones.
\iteme{PARI(66) :} sum of the transverse momenta of all partons
generated at the hardest interaction, including initial- and
final-state radiation, resonance decay products, and primordial
$k_{\perp}$.
\iteme{PARI(67) :} scalar sum of transverse momenta of partons generated
at hard interactions, excluding the hardest one, along with its
initial- and final-state radiation (see \ttt{PARI(66)}). Is
non-vanishing only in the multiple-interaction scenarios. In the
new scenario the initial- and final-state radiation associated with
further interactions is included.
\iteme{PARI(68) :} currently equal to \ttt{PARI(67)}.
\iteme{PARI(69) :} sum of transverse momenta of all partons generated
in hard interactions (\ttt{PARI(66) + PARI(68)}) and, additionally,
of all beam-remnant partons.
\iteme{PARI(71), PARI(72) :} sum of the momentum fractions $x$ taken
by initial-state parton-shower initiators on side 1 and and side 2,
excluding those of the hardest interaction. Is non-vanishing only in
the multiple-interaction scenario.
\iteme{PARI(73), PARI(74) :} sum of the momentum fractions $x$ taken
by the partons at the hard interaction on side 1 and side 2, excluding
those of the hardest interaction. Is non-vanishing only in the
multiple-interaction scenario.
\iteme{PARI(75), PARI(76) :} the $x$ value of a photon that branches
into quarks or gluons, i.e.\ $x$ at interface between initial-state QED
and QCD cascades, for the old photoproduction machinery.
\iteme{PARI(77), PARI(78) :} the $\chi$ values selected for beam
remnants that are split into two objects, describing how the energy
is shared (see \ttt{MSTP(92)} and \ttt{MSTP(94)}); is vanishing if
no splitting is needed.
\iteme{PARI(81) :} size of the threshold factor (enhancement or
suppression) in the latest event with heavy-flavour production;
see \ttt{MSTP(35)}.
\iteme{PARI(91) :} average multiplicity $\br{n}$ of pile-up events,
see \ttt{MSTP(133)}. Only relevant for \ttt{MSTP(133) = } 1 or 2.
\iteme{PARI(92) :} average multiplicity $\langle n \rangle$ of pile-up
events as actually simulated, i.e.\ with multiplicity = 0 events
removed and the high-end tail truncated. Only relevant for
\ttt{MSTP(133) =} 1 or 2.
\iteme{PARI(93) :} for \ttt{MSTP(133) = 1} it is the probability that
a beam crossing will produce a pile-up event at all, i.e.\ that there
will be at least one hadron--hadron interaction; for
\ttt{MSTP(133) = 2} the probability that a beam crossing will produce
a pile-up event with one hadron--hadron interaction of the desired rare
type. See section \ref{ss:pileup}.
\iteme{PARI(101) :} c.m.\ energy for the full collision, while \ttt{PARI(11)}
gives the $\gamma$-hadron or $\gamma\gamma$ subsystem energy; used for
virtual photons generated internally with the {\galep} option.
\iteme{PARI(102) :} full squared c.m.\ energy, while \ttt{PARI(12)} gives
the subsystem squared energy; used for virtual photons generated
internally with the {\galep} option.
\iteme{PARI(103), PARI(104) :} $x$ values, i.e.\ respective photon energy
fractions of the incoming lepton in the c.m.\ frame of the event; used for
virtual photons generated internally with the {\galep} option.
\iteme{PARI(105), PARI(106) :} $Q^2$ or $P^2$, virtuality of the
respective photon (thus the square of \ttt{VINT(3)}, \ttt{VINT(4)}); used for
virtual photons generated internally with the {\galep} option.
\iteme{PARI(107), PARI(108) :} $y$ values, i.e.\ respective photon light-cone
energy fraction of the incoming lepton; used for virtual photons generated
internally with the {\galep} option.
\iteme{PARI(109), PARI(110) :} $\theta$, scattering angle of the respective
lepton in the c.m.\ frame of the event; used for virtual photons generated
internally with the {\galep} option.
\iteme{PARI(111), PARI(112) :} $\phi$, azimuthal angle of the respective
scattered lepton in the c.m.\ frame of the event; used for virtual photons
generated internally with the {\galep} option.
\iteme{PARI(113), PARI(114):} the $R$ factor defined at \ttt{MSTP(17)},
giving a cross section enhancement from the contribution of resolved
longitudinal photons.
\end{entry}
\subsection{How to Generate Weighted Events}
\label{ss:PYwtevts}
By default {\Py} generates unweighted events, i.e.\ all events
in a run are on an equal footing. This means that corners of phase
space with low cross sections are poorly populated, as it should be.
However, sometimes one is interested in also exploring such corners,
in order to gain a better understanding of physics. A typical example
would be the jet cross section in hadron collisions, which is dropping
rapidly with increasing jet $\pT$, and where it is interesting to trace
this drop over several orders of magnitude. Experimentally this may
be solved by prescaling event rates already at the trigger level,
so that all high-$\pT$ events are saved but only a fraction of the
lower-$\pT$ ones. In this section we outline procedures to generate
events in a similar manner.
Basically two approaches can be used. One is to piece together
results from different subruns, where each subrun is restricted to
some specific region of phase space. Within each subrun all events
then have the same weight, but subruns have to be combined according
to their relative cross sections. The other approach is to let each
event come with an associated weight, that can vary smoothly as a
function of $\pT$. These two alternatives correspond to stepwise or
smoothly varying prescaling factors in the experimental analogue.
We describe them one after the other.
The phase space can be sliced in many different ways. However, for the
jet rate and many other processes, the most natural variable would be
$\pT$ itself. (For production of a lepton pair by $s$-channel
resonances, the invariant mass would be a better choice.) It is not
possible to specify beforehand the jet $\pT$'s an event will contain,
since this is a combination of the $\hat{p}_{\perp}$ of the hard
scattering process with additional showering activity, with
hadronization, with underlying event and with the jet clustering
approach actually used. However, one would expect a strong correlation
between the $\hat{p}_{\perp}$ scale and the jet $\pT$'s. Therefore
the full $\hat{p}_{\perp}$ range can be subdivided into a set of
ranges by using the \ttt{CKIN(3)} and \ttt{CKIN(4)} variables as
lower and upper limits. This could be done e.g.\ for adjacent
non-overlapping bins 10--20,20--40,40--70, etc.
Only if one would like to cover also very small $\pT$ is there a
problem with this strategy: since the na\"{\i}ve jet cross section is
divergent for $\hat{p}_{\perp} \to 0$, a unitarization procedure is
implied by setting \ttt{CKIN(3) = 0} (or some other low value). This
unitarization then disregards the actual \ttt{CKIN(3)} and \ttt{CKIN(4)}
values and generates events over the full phase space. In order not to
double-count, then events above the intended upper limit of the first
bin have to be removed by brute force.
A simple but complete example of a code performing this task (with
some primitive histogramming) is the following:
\begin{verbatim}
C...All real arithmetic in double precision.
IMPLICIT DOUBLE PRECISION(A-H, O-Z)
C...Three Pythia functions return integers, so need declaring.
INTEGER PYK,PYCHGE,PYCOMP
C...EXTERNAL statement links PYDATA on most platforms.
EXTERNAL PYDATA
C...The event record.
COMMON/PYJETS/N,NPAD,K(4000,5),P(4000,5),V(4000,5)
C...Selection of hard-scattering subprocesses.
COMMON/PYSUBS/MSEL,MSELPD,MSUB(500),KFIN(2,-40:40),CKIN(200)
C...Parameters.
COMMON/PYPARS/MSTP(200),PARP(200),MSTI(200),PARI(200)
C...Bins of pT.
DIMENSION PTBIN(10)
DATA PTBIN/0D0,10D0,20D0,40D0,70D0,110D0,170D0,250D0,350D0,1000D0/
C...Main parameters of run: c.m. energy and number of events per bin.
ECM=2000D0
NEV=1000
C...Histograms.
CALL PYBOOK(1,'dn_ev/dpThat',100,0D0,500D0)
CALL PYBOOK(2,'dsigma/dpThat',100,0D0,500D0)
CALL PYBOOK(3,'log10(dsigma/dpThat)',100,0D0,500D0)
CALL PYBOOK(4,'dsigma/dpTjet',100,0D0,500D0)
CALL PYBOOK(5,'log10(dsigma/dpTjet)',100,0D0,500D0)
CALL PYBOOK(11,'dn_ev/dpThat, dummy',100,0D0,500D0)
CALL PYBOOK(12,'dn/dpTjet, dummy',100,0D0,500D0)
C...Loop over pT bins and initialize.
DO 300 IBIN=1,9
CKIN(3)=PTBIN(IBIN)
CKIN(4)=PTBIN(IBIN+1)
CALL PYINIT('CMS','p','pbar',ECM)
C...Loop over events. Remove unwanted ones in first pT bin.
DO 200 IEV=1,NEV
CALL PYEVNT
PTHAT=PARI(17)
IF(IBIN.EQ.1.AND.PTHAT.GT.PTBIN(IBIN+1)) GOTO 200
C...Store pThat. Cluster jets and store variable number of pTjet.
CALL PYFILL(1,PTHAT,1D0)
CALL PYFILL(11,PTHAT,1D0)
CALL PYCELL(NJET)
DO 100 IJET=1,NJET
CALL PYFILL(12,P(N+IJET,5),1D0)
100 CONTINUE
C...End of event loop.
200 CONTINUE
C...Normalize cross section to pb/GeV and add up.
FAC=1D9*PARI(1)/(DBLE(NEV)*5D0)
CALL PYOPER(2,'+',11,2,1D0,FAC)
CALL PYOPER(4,'+',12,4,1D0,FAC)
C...End of loop over pT bins.
300 CONTINUE
C...Take logarithm and plot.
CALL PYOPER(2,'L',2,3,1D0,0D0)
CALL PYOPER(4,'L',4,5,1D0,0D0)
CALL PYNULL(11)
CALL PYNULL(12)
CALL PYHIST
END
\end{verbatim}
The alternative to slicing the phase space is to used weighted events.
This is possible by making use of the \ttt{PYEVWT} routine:
\drawbox{CALL PYEVWT(WTXS)}\label{p:PYEVWT}
\begin{entry}
\itemc{Purpose:} to allow you to reweight event cross sections,
by process type and kinematics of the hard scattering. There exists
two separate modes of usage, described in the following. \\
For \ttt{MSTP(142) = 1}, it is assumed that the cross section of the
process is correctly given by default in {\Py}, but that one
wishes to generate events biased to a specific region of phase
space. While the \ttt{WTXS} factor therefore multiplies the na\"{\i}ve
cross section in the choice of subprocess type and kinematics,
the produced event comes with a compensating weight
\ttt{PARI(10) = 1./WTXS}, which should be used when filling histograms
etc. In the \ttt{PYSTAT(1)} table, the cross sections are unchanged
(up to statistical errors) compared with the standard cross sections,
but the relative composition of events may be changed and need
no longer be in proportion to relative cross sections. A typical
example of this usage is if one wishes to enhance the production
of high-$\pT$ events; then a weight like
\ttt{WTXS}$=(\pT/\pTzero)^2$ (with $\pTzero$ some fixed
number) might be appropriate. See \ttt{PARI(2)} for a discussion of
overall normalization issues.\\
For \ttt{MSTP(142) = 2}, on the other hand, it is assumed that the true
cross section is really to be modified by the multiplicative factor
\ttt{WTXS}. The generated events therefore come with unit weight, just
as usual. This option is really equivalent to replacing the basic
cross sections coded in {\Py}, but allows more flexibility: no
need to recompile the whole of {\Py}. \\
The routine will not be called unless \ttt{MSTP(142)} $\geq 1$, and
never if `minimum-bias'-type events (including elastic and
diffractive scattering) are to be generated as well. Further,
cross sections for additional multiple interactions or pile-up events
are never affected. A dummy routine \ttt{PYEVWT} is included in the
program file, so as to avoid unresolved external references when the
routine is not used.
\iteme{WTXS:} multiplication factor to ordinary event cross section;
to be set (by you) in \ttt{PYEVWT} call.
\itemc{Remark :} at the time of selection, several variables in the
\ttt{MINT} and \ttt{VINT} arrays in the \ttt{PYINT1} common block
contain information that can be used to make the decision. The routine
provided in the program file explicitly reads the variables that
have been defined at the time \ttt{PYEVWT} is called, and also
calculates some derived quantities. The given list of information
includes subprocess type {\ISUB}, $E_{\mrm{cm}}$, $\hat{s}$, $\hat{t}$,
$\hat{u}$, $\hat{p}_{\perp}$, $x_1$, $x_2$, $x_{\mrm{F}}$, $\tau$, $y$,
$\tau'$, $\cos\hat{\theta}$, and a few more. Some of these may
not be relevant for the process under study, and are then set to
zero.
\itemc{Warning:} the weights only apply to the hard-scattering
subprocesses. There is no way to reweight the shape of initial- and
final-state showers, fragmentation, or other aspects of the event.
\end{entry}
\boxsep
There are some limitations to the facility. \ttt{PYEVWT} is called
at an early stage of the generation process, when the hard kinematics
is selected, well before the full event is constructed. It then cannot
be used for low-$\pT$, elastic or diffractive events, for which no
hard kinematics has been defined. If such processes are included,
the event weighting is switched off. Therefore it is no longer an
option to run with \ttt{CKIN(3) = 0}.
Which weight expression to use may take some trial and error.
In the above case, a reasonable ansatz seems to be a weight behaving
like $\hat{p}_{\perp}^6$, where four powers of $\hat{p}_{\perp}$ are
motivated by the partonic cross section behaving like
$1/\hat{p}_{\perp}^4$, and the remaining two by the fall-off of
parton densities. An example for the same task as above one would
then be:
\begin{verbatim}
C...All real arithmetic in double precision.
IMPLICIT DOUBLE PRECISION(A-H, O-Z)
C...Three Pythia functions return integers, so need declaring.
INTEGER PYK,PYCHGE,PYCOMP
C...EXTERNAL statement links PYDATA on most platforms.
EXTERNAL PYDATA
C...The event record.
COMMON/PYJETS/N,NPAD,K(4000,5),P(4000,5),V(4000,5)
C...Selection of hard-scattering subprocesses.
COMMON/PYSUBS/MSEL,MSELPD,MSUB(500),KFIN(2,-40:40),CKIN(200)
C...Parameters.
COMMON/PYPARS/MSTP(200),PARP(200),MSTI(200),PARI(200)
C...Main parameters of run: c.m. energy, pTmin and number of events.
ECM=2000D0
CKIN(3)=5D0
NEV=10000
C...Histograms.
CALL PYBOOK(1,'dn_ev/dpThat',100,0D0,500D0)
CALL PYBOOK(2,'dsigma/dpThat',100,0D0,500D0)
CALL PYBOOK(3,'log10(dsigma/dpThat)',100,0D0,500D0)
CALL PYBOOK(4,'dsigma/dpTjet',100,0D0,500D0)
CALL PYBOOK(5,'log10(dsigma/dpTjet)',100,0D0,500D0)
C...Initialize with weighted events.
MSTP(142)=1
CALL PYINIT('CMS','p','pbar',ECM)
C...Loop over events; read out pThat and event weight.
DO 200 IEV=1,NEV
CALL PYEVNT
PTHAT=PARI(17)
WT=PARI(10)
C...Store pThat. Cluster jets and store variable number of pTjet.
CALL PYFILL(1,PTHAT,1D0)
CALL PYFILL(2,PTHAT,WT)
CALL PYCELL(NJET)
DO 100 IJET=1,NJET
CALL PYFILL(4,P(N+IJET,5),WT)
100 CONTINUE
C...End of event loop.
200 CONTINUE
C...Normalize cross section to pb/GeV, take logarithm and plot.
FAC=1D9*PARI(2)/5D0
CALL PYFACT(2,FAC)
CALL PYFACT(4,FAC)
CALL PYOPER(2,'L',2,3,1D0,0D0)
CALL PYOPER(4,'L',4,5,1D0,0D0)
CALL PYHIST
END
C*************************************************************
SUBROUTINE PYEVWT(WTXS)
C...Double precision and integer declarations.
IMPLICIT DOUBLE PRECISION(A-H, O-Z)
IMPLICIT INTEGER(I-N)
INTEGER PYK,PYCHGE,PYCOMP
C...Common block.
COMMON/PYINT1/MINT(400),VINT(400)
C...Read out pThat^2 and set weight.
PT2=VINT(48)
WTXS=PT2**3
RETURN
END
\end{verbatim}
Note that, in \ttt{PYEVWT} one cannot look for $\hat{p}_{\perp}$ in
\ttt{PARI(17)}, since this variable is only set at the end of the
event generation. Instead the internal \ttt{VINT(48)} is used.
The dummy copy of the \ttt{PYEVWT} routine found in the {\Py} code
shows what is available and how to access this.
\boxsep
The above solutions do not work if the weighting should also depend
on the shower evolution, i.e.\ not only on the hard process. Such a
rejection/acceptance `weight' is particularly convenient when
trying to combine matrix-element-generated events of different jet
multiplicities (in addition to whatever is the basic process) without
doublecounting, e.g.\ using the L--CKKW or MLM matching prescriptions
\cite{Cat01}. To handle such situations, a routine \ttt{UPVETO} has
been introduced. The name is intended to emphasize the logical place
along with the \ttt{UPINIT} and \ttt{UPEVNT} routines for the handling
of external user processes (see section \ref{ss:PYnewproc}), but it
can also be used for internal {\Py} processes.
\drawbox{CALL UPVETO(IVETO)}\label{p:UPVETO}
\begin{entry}
\itemc{Purpose:} to allow the user the possibility to abort the
generation of an event immediately after parton showers but before
underlying event and handronization is considered. The routine is called
only if \ttt{MSTP(143) = 1}, and only for the old, virtuality-ordered
showers in \ttt{PYEVNT}, i.e.\ has not been implemented for the
$\pT$-ordered showers in \ttt{PYEVNW}.
\iteme{IVETO :} specifies choice made by user.
\begin{subentry}
\iteme{= 0 :} retain current event and generate it in full.
\iteme{= 1 :} abort generation of current event and move to next.
\end{subentry}
\itemc{Note 1:} if resonances like $\W^{\pm}$, $\Z^0$, $\t$, Higgs and
SUSY particles are handed undecayed from \ttt{UPEVNT}, or are generated
internally in {\Py}, they will also be undecayed at this stage; if decayed
their decay products will have been allowed to shower.
\itemc{Note 2:} all partons at the end of the shower phase are stored in
the \ttt{HEPEVT} commonblock, see section \ref{ss:HEPEVT}.
The interesting information is:
\begin{subentry}
\iteme{NHEP :} the number of such partons, in entries 1 through \ttt{NHEP},
\iteme{ISTHEP(I) :} where all intermediate products have code 2 and
all the final ones (at the time \ttt{UPVETO} is called) code 1,
\iteme{IDHEP(I) :} the particle identity code according to PDG conventions,
\iteme{JMOHEP(1,I) :} the mother position,
\iteme{PHEP(J,I) :} the $(p_x, p_y, p_z, E, m)$ of the particle.
\end{subentry}
The rest of the \ttt{HEPEVT} variables are zeroed.
\itemc{Note 3:} the cross section of processes, as shown with
\ttt{CALL PYSTAT(1)}, is reduced when events are aborted. Such events
are also counted in the "fraction of events that fail fragmentation cuts"
in the last line of this table.
\end{entry}
\boxsep
In the \ttt{HEPEVT} event record, all intermediate and `final'
particles of the hard process itself, i.e.\ of the matrix-element
calculation, are listed as documentation lines, with \ttt{ISTHEP(I) = 2}.
The `final'-state particles that actually are defined when \ttt{UPVETO}
is called, after shower evolution but before multiple interactions have
been added, have \ttt{ISTHEP(I) = 1}. These point back to the one of the
\ttt{ISTHEP(I) = 2} partons they originate from, as first mother. If a
system does not radiate, the same set of partons will be repeated twice,
once with \ttt{ISTHEP(I) = 2} and once with \ttt{ISTHEP(I) = 1}. A more
typical example would be that a set of partons with \ttt{ISTHEP(I) = 1}
point back to the same `mother' with \ttt{ISTHEP(I) = 2}. Note that
all the intermediate stages of the shower evolution are not shown,
only the original mother and all the final daughters. If the
mother index is zero for an \ttt{ISTHEP(I) = 1} parton, it comes
from initial-state radiation.
Of course, the separation of which radiation comes from where is often
gauge-dependent, i.e.\ model-dependent, and so should not be over-stressed.
For instance, in a $\t \to \b \W^+ \to \b \u \dbar$ decay sequence, the
gluon radiation off the $\b$ is rather unambiguous, while $\u$ and $\dbar$
form part of the same radiating system, although some sensible separation
is provided by the program.
\subsection{How to Run with Varying Energies}
\label{ss:PYvaren}
It is possible to use {\Py} in a mode where the energy can be
varied from one event to the next, without the need to re-initialize
with a new \ttt{PYINIT} call. This allows a significant speed-up of
execution, although it is not as fast as running at a fixed
energy. It can not be used for everything --- we will come to
the fine print at the end --- but it should be applicable for most
tasks.
The master switch to access this possibility is in \ttt{MSTP(171)}.
By default it is off, so you must set \ttt{MSTP(171) = 1} before
initialization. There are two submodes of running, with
\ttt{MSTP(172)} being 1 or 2. In the former mode, {\Py} will generate
an event at the requested energy. This means that you have to know
which energy you want beforehand. In the latter mode, {\Py}
will often return without having generated an event --- with flag
\ttt{MSTI(61) = 1} to signal that --- and you are then requested to give
a new energy. The energy spectrum of accepted events will then,
in the end, be your na\"{\i}ve input spectrum weighted with the
cross-section of the processes you study. We will come back to
this.
The energy can be varied, whichever frame is given in the \ttt{PYINIT}
call. (Except for \ttt{'USER'}, where such information is fed in via
the \ttt{HEPEUP} common block and thus beyond the control of {\Py}.)
When the frame is \ttt{'CMS'}, \ttt{PARP(171)} should be filled
with the fractional energy of each event, i.e.\
$E_{\mrm{cm}} =$\ttt{PARP(171)}$\times$\ttt{WIN}, where \ttt{WIN} is
the nominal c.m.\ energy of the \ttt{PYINIT} call. Here \ttt{PARP(171)}
should normally be smaller than unity, i.e.\ initialization should
be done at the maximum energy to be encountered. For the \ttt{'FIXT'}
frame, \ttt{PARP(171)} should be filled by the fractional beam energy
of that one, i.e.\ $E_{\mrm{beam}} = $\ttt{PARP(171)}$\times$\ttt{WIN}.
For the \ttt{'3MOM'}, \ttt{'4MOM'} and \ttt{'5MOM'} options, the
two four-momenta are given in for each event in the same format as
used for the \ttt{PYINIT} call. Note that there is a minimum c.m.\
energy allowed, \ttt{PARP(2)}. If you give in values below this,
the program will stop for \ttt{MSTP(172) = 1}, and will return with
\ttt{MSTI(61) = 1} for \ttt{MSTP(172) = 1}.
To illustrate the use of the \ttt{MSTP(172) = 2} facility, consider the
case of beamstrahlung in $\ee$ linear colliders. This is just for
convenience; what is said here can be translated easily into other
situations. Assume that the beam spectrum is given by $D(z)$, where
$z$ is the fraction retained by the original $\e$ after beamstrahlung.
Therefore $0 \leq z \leq 1$ and the integral of $D(z)$ is unity.
This is not perfectly general; one could imagine branchings
$\e^- \to \e^- \gamma \to \e^-\e^+\e^-$, which gives a multiplication
in the number of beam particles. This could either be expressed in
terms of a $D(z)$ with integral larger than unity or in terms of an
increased luminosity. We will assume the latter, and use $D(z)$
properly normalized. Given a nominal $s = 4E_{\mrm{beam}}^2$,
the actual $s'$ after beamstrahlung is given by $s' = z_1 z_2 s$.
For a process with a cross section $\sigma(s)$ the total
cross section is then
\begin{equation}
\sigma_{\mrm{tot}} = \int_0^1 \int_0^1 D(z_1) \, D(z_2)
\sigma(z_1 z_2 s) \, \d z_1 \, \d z_2~.
\end{equation}
The cross section $\sigma$ may in itself be an integral over a number
of additional phase space variables. If the maximum of the differential
cross section is known, a correct procedure to generate events is
\begin{Enumerate}
\item pick $z_1$ and $z_2$ according to $D(z_1) \, \d z_1$ and
$D(z_2) \, \d z_2$, respectively;
\item pick a set of phase space variables of the process, for the
given $s'$ of the event;
\item evaluate $\sigma(s')$ and compare with $\sigma_{\mmax}$;
\item if event is rejected, then return to step 1 to generate new
variables;
\item else continue the generation to give a complete event.
\end{Enumerate}
You as a user are assumed to take care of step 1, and present the
resulting kinematics with incoming $\e^+$ and $\e^-$ of varying energy.
Thereafter {\Py} will do steps 2--5, and either return an event or
put \ttt{MSTI(61) = 1} to signal failure in step 4.
The maximization procedure does search in phase space to find
$\sigma_{\mmax}$, but it does not vary the $s'$ energy in this process.
Therefore the maximum search in the \ttt{PYINIT} call should be
performed where the cross section is largest. For processes with
increasing cross section as a function of energy this means at the
largest energy that will ever be encountered, i.e.\ $s' = s$ in the
case above. This is the `standard' case, but often one encounters
other behaviours, where more complicated procedures are needed. One
such case would be the process $\ee \to \Z^{*0} \to \Z^0 \hrm^0$,
which is known to have a cross section that increases near the
threshold but is decreasing asymptotically. If one already knows
that the maximum, for a given Higgs mass, appears at 300 GeV, say,
then the \ttt{PYINIT} call should be made with that energy, even if
subsequently one will be generating events for a 500 GeV collider.
In general, it may be necessary to modify the selection of $z_1$ and
$z_2$ and assign a compensating event weight. For instance, consider
a process with a cross section behaving roughly like $1/s$. Then the
$\sigma_{\mrm{tot}}$ expression above may be rewritten as
\begin{equation}
\sigma_{\mrm{tot}} = \int_0^1 \int_0^1 \frac{D(z_1)}{z_1} \,
\frac{D(z_2)}{z_2} \, z_1 z_2 \sigma(z_1 z_2 s) \, \d z_1 \, \d z_2 ~.
\end{equation}
The expression $z_1 z_2 \sigma(s')$ is now essentially flat in $s'$,
i.e.\ not only can $\sigma_{\mmax}$ be found at a convenient energy
such as the maximum one, but additionally the {\Py} generation
efficiency (the likelihood of surviving step 4) is greatly
enhanced. The price to be paid is that $z$ has to be selected
according to $D(z)/z$ rather than according to $D(z)$. Note that
$D(z)/z$ is not normalized to unity. One therefore needs to define
\begin{equation}
{\cal I}_D = \int_0^1 \frac{D(z)}{z} \, \d z ~,
\end{equation}
and a properly normalized
\begin{equation}
D'(z) = \frac{1}{{\cal I}_D} \, \frac{D(z)}{z} ~.
\end{equation}
Then
\begin{equation}
\sigma_{\mrm{tot}} = \int_0^1 \int_0^1 D'(z_1) \, D'(z_2) \,
{\cal I}_D^2 \, z_1 z_2 \sigma(z_1 z_2 s) \, \d z_1 \, \d z_2 ~.
\end{equation}
Therefore the proper event weight is ${\cal I}_D^2 \, z_1 z_2$.
This weight should be stored by you, for each event, in \ttt{PARP(173)}.
The maximum weight that will be encountered should be stored in
\ttt{PARP(174)} before the \ttt{PYINIT} call, and not changed
afterwards. It is not necessary to know the precise maximum;
any value larger than the true maximum will do, but the inefficiency
will be larger the cruder the approximation.
Additionally you must put \ttt{MSTP(173) = 1} for the program to
make use of weights at all. Often $D(z)$ is not known analytically;
therefore ${\cal I}_D$ is also not known beforehand, but may have to
be evaluated (by you) during the course of the run. Then you should
just use the weight $z_1 z_2$ in \ttt{PARP(173)} and do the overall
normalization yourself in the end. Since \ttt{PARP(174) = 1} by
default, in this case you need not set this variable specially.
Only the cross sections are affected by the procedure selected for
overall normalization, the events themselves still are properly
distributed in $s'$ and internal phase space.
Above it has been assumed tacitly that $D(z) \to 0$ for $z \to 0$.
If not, $D(z)/z$ is divergent, and it is not possible to define a
properly normalized $D'(z) = D(z)/z$. If the cross section is truly
diverging like $1/s$, then a $D(z)$ which is nonvanishing for
$z \to 0$ does imply an infinite total cross section, whichever way
things are considered. In cases like that, it is necessary to impose
a lower cut on $z$, based on some physics or detector consideration.
Some such cut is anyway needed to keep away from the minimum c.m.\
energy required for {\Py} events, see above.
The most difficult cases are those with a very narrow and high
peak, such as the $\Z^0$. One could initialize at the energy of
maximum cross section and use $D(z)$ as is, but efficiency might
turn out to be very low. One might then be tempted to do more
complicated transforms of the kind illustrated above. As a rule
it is then convenient to work in the variables $\tau_z = z_1 z_2$
and $y_z = (1/2) \ln (z_1/z_2)$, cf. section \ref{ss:kinemtwo}.
Clearly, the better the behaviour of the cross section can be
modelled in the choice of $z_1$ and $z_2$, the better the overall
event generation efficiency. Even under the best of circumstances,
the efficiency will still be lower than for runs with fix energy.
There is also a non-negligible time overhead for using variable
energies in the first place, from kinematics reconstruction
and (in part) from the phase space selection. One should therefore
not use variable energies when not needed, and not use a large
range of energies $\sqrt{s'}$ if in the end only a smaller range
is of experimental interest.
This facility may be combined with most other aspects of the program.
For instance, it is possible to simulate beamstrahlung as above and
still include bremsstrahlung with \ttt{MSTP(11) = 1}. Further, one may
multiply the overall event weight of \ttt{PARP(173)} with a
kinematics-dependent weight given by \ttt{PYEVWT}, although it is not
recommended (since the chances of making a mistake are also
multiplied). However, a few things do \textit{not} work.
\begin{Itemize}
\item It is not possible to use pile-up events, i.e.\ you must have
\ttt{MSTP(131) = 0}.
\item The possibility of giving in your own cross-section optimization
coefficients, option \ttt{MSTP(121) = 2}, would require more input
than with fixed energies, and this option should therefore not be
used. You can still use \ttt{MSTP(121) = 1}, however.
\item The multiple interactions scenario with \ttt{MSTP(82)} $\geq 2$
only works approximately for energies different from the initialization
one. If the c.m.\ energy spread is smaller than a factor 2, say, the
approximation should be reasonable, but if the spread is larger
one may have to subdivide into subruns of different energy bins.
The initialization should be made at the largest energy to be
encountered --- whenever multiple interactions are possible (i.e.\ for
incoming hadrons and resolved photons) this is where the cross sections
are largest anyway, and so this is no further constraint. There is no
simple possibility to change \ttt{PARP(82)} during the course of the
run, i.e.\ an energy-independent $\pTzero$ must be assumed. By contrast,
for \ttt{MSTP(82) = 1} $\pTmin=$\ttt{PARP(81)} can be set differently
for each event, as a function of c.m.\ energy. Initialization should
then be done with \ttt{PARP(81)} as low as it is ever supposed to become.
\end{Itemize}
\subsection{How to Include External Processes}
\label{ss:PYnewproc}
Despite a large repertory of processes in {\Py}, the number of
interesting missing ones clearly is even larger, and with time this
discrepancy is likely to increase. There are several reasons why it is
not practicable to imagine a {\Py} which has `everything'. One is
the amount of time it takes to implement a process for the few
{\Py} authors, compared with the rate of new cross section results
produced by the rather larger matrix-element calculations community.
Another is the length of currently produced matrix-element expressions,
which would make the program very bulky. A third argument is that,
whereas the phase space of $2 \to 1$ and $2 \to 2$ processes can be set
up once and for all according to a reasonably flexible machinery,
processes with more final-state particles are less easy to generate.
To achieve a reasonable efficiency, it is necessary to tailor the
phase-space selection procedure to the dynamics of the given process,
and to the desired experimental cuts.
At times, simple solutions may be found. Some processes may be seen just
as trivial modifications of already existing ones. For instance, you
might want to add some extra term, corresponding to contact interactions,
to the matrix elements of a {\Py} $2 \to 2$ process. In that case it is
not necessary to go through the machinery below, but instead you can use
the \ttt{PYEVWT} routine (section \ref{ss:PYwtevts}) to introduce an
additional weight for the event, defined as the ratio of the modified to
the unmodified differential cross sections. If you use the option
\ttt{MSTP(142) = 2}, this weight is considered as part of the `true'
cross section of the process, and the generation is changed accordingly.
A {\Py} expert could also consider implementing a new process along the
lines of the existing ones, hardwired in the code. Such a modification
would have to be ported anytime the {\Py} program is upgraded, however
(unless it is made available to the {\Py} authors and incorporated into
the public distribution). For this and other reasons, we will not
consider this option in detail, but only provide a few generic remarks.
The first step is to pick a process number \ttt{ISUB} among ones not in
use. The process type needs to be set in \ttt{ISET(ISUB)} and, if the
final state consists of massive particles, these should be specified in
\ttt{KFPR(ISUB,1)} and \ttt{KFPR(ISUB,2)}. Output is improved if a
process name is set in \ttt{PROC(ISUB)}. The second and main step is to
code the cross section of the hard-scattering subprocess in the
\ttt{PYSIGH} routine. Usually the best starting point is to use the code
of an existing similar process as a template for the new code required.
The third step is to program the selection of the final state in
\ttt{PYSCAT}, normally a simple task, especially if again a similar
process (especially with respect to colour flow) can be used as template.
In many cases the steps above are enough, in others additional
modifications are required to \ttt{PYRESD} to handle process-specific
non-isotropic decays of resonances. Further code may also be required
e.g. if a process can proceed via an intermediate resonance that can be
on the mass shell.
The recommended solution, if a desired process is missing, is instead
to include it into {\Py} as an `external' process. In this section we
will describe how it is possible to specify the parton-level state of some
hard-scattering process in a common block. (`Parton-level' is not intended
to imply a restriction to quarks and gluons as interacting particles,
but only that quarks and gluons are given rather than the hadrons they will
produce in the observable final state.) {\Py} will read this common
block, and add initial- and final-state showers, beam remnants and
underlying events, fragmentation and decays, to build up an event in as much
detail as an ordinary {\Py} one. Another common block is to be filled with
information relevant for the run as a whole, where beams and processes
are specified.
%Such a facility has been available since long, and has been used e.g.\
%referee: misprint
Such a facility has been available since long ago, and has been used e.g.\
together with the \tsc{CompHEP} package. \tsc{CompHEP} \cite{Puk99} is
mainly intended for the automatic computation of matrix elements, but also
allows the sampling of phase space according to these matrix elements
and thereby the generation of weighted or unweighted events. These
events can be saved on disk and thereafter read back in to {\Py} for
subsequent consideration \cite{Bel00}.
At the Les Houches 2001 workshop it was decided to develop a common
standard, that could be used by all matrix-elements-based generators to
feed information into any complete event generator --- the Les Houches
Accord (LHA) \cite{Boo01}. It is
similar to, but in its details different from, the approach previously
implemented in {\Py}. Furthermore, it uses the same naming convention:
all names in common blocks end with \ttt{UP}, short for User(-defined)
Process. This produces some clashes. Therefore the old facility,
existing up to and including {\Py}~6.1, has been completely removed and
replaced by the new one. Currently not every last detail of the standard
has yet been implemented. In the description below we will emphasize
such restrictions, as well as the solutions to aspects not specified
by the standard.
In particular, even with the common-block contents defined, it is not
clear where they are to be filled, i.e.\ how the external supplier of
parton-level events should synchronize with {\Py}. The solution adopted
here --- recommended in the standard --- is to introduce two subroutines,
\ttt{UPINIT} and \ttt{UPEVNT}. The first is called by \ttt{PYINIT} at
initialization to obtain information about the run itself, and the other
called by \ttt{PYEVNT} each time a new event configuration is to be fed
in. We begin by describing these two steps and their related common blocks,
before proceeding with further details and examples. The description is
cast in a {\Py}-oriented language, but for the common-block contents it
closely matches the generator-neutral standard in \cite{Boo01}.
Restrictions to or extensions of the standard should be easily recognized,
but in case you are vitally dependent on following the standard exactly,
you should of course check \cite{Boo01}.
Note that the \ttt{UPVETO} routine, introduced above in section
\ref{ss:PYwtevts}, is a third member of the \ttt{UP....} family of
routines. While not part of the LHA, it offers extra
functionality that allows the user to combine parton configuration input
at different orders of perturbation theory, in such a way that the
matrix-elements description and the shower activity do not
doublecount emissions.
\subsubsection{Run information}
When \ttt{PYINIT} is called in the main program, with \ttt{'USER'} as first
argument (which makes the other arguments dummy), it signals that external
processes are to be implemented. Then \ttt{PYINIT}, as part of its
initialization tasks, will call the routine \ttt{UPINIT}.
\drawbox{CALL UPINIT}\label{p:UPINIT}
\begin{entry}
\itemc{Purpose:} routine to be provided by you when you want to implement
external processes, wherein the contents of the \ttt{HEPRUP} common block
are set. This information specifies the character of the run, both beams and
processes, see further below.
\itemc{Note 1:} alternatively, the \ttt{HEPRUP} common block could be filled
already before \ttt{PYINIT} is called, in which case \ttt{UPINIT} could be
empty. We recommend \ttt{UPINIT} as the logical place to collect the relevant
information, however.
\itemc{Note 2:} a dummy copy of \ttt{UPINIT} is distributed with the
program, in order to avoid potential problems with unresolved external
references. This dummy should not be linked when you supply your own
\ttt{UPINIT} routine. The code can be used to read initialization
information previously written by \ttt{PYUPIN}, however.
\end{entry}
\drawboxseven{~INTEGER MAXPUP}%
{~PARAMETER (MAXPUP=100)}%
{~INTEGER IDBMUP,PDFGUP,PDFSUP,IDWTUP,NPRUP,LPRUP}%
{~DOUBLE PRECISION EBMUP,XSECUP,XERRUP,XMAXUP}%
{~COMMON/HEPRUP/IDBMUP(2),EBMUP(2),PDFGUP(2),PDFSUP(2),}%
{\&IDWTUP,NPRUP,XSECUP(MAXPUP),XERRUP(MAXPUP),XMAXUP(MAXPUP),}%
{\&LPRUP(MAXPUP)}\label{p:HEPRUP}
\begin{entry}
\itemc{Purpose:} to contain the initial information necessary for the
subsequent generation of complete events from externally provided parton
configurations. The \ttt{IDBMUP}, \ttt{EBMUP}, \ttt{PDFGUP} and
\ttt{PDFSUP} variables specify the nature of the two incoming beams.
\ttt{IDWTUP} is a master switch, selecting the strategy to be used to mix
different processes. \ttt{NPRUP} gives the number of different external
processes to mix, and \ttt{XSECUP}, \ttt{XERRUP}, \ttt{XMAXUP} and
\ttt{LPRUP} information on each of these. The contents in this common block
must remain unchanged by the user during the course of the run, once set
in the initialization stage.\\
This common block should be filled in the \ttt{UPINIT} routine or,
alternatively, before the \ttt{PYINIT} call. During the run, {\Py} may
update the \ttt{XMAXUP} values as required.
\iteme{MAXPUP :}\label{p:MAXPUP} the maximum number of distinguishable
processes that can be defined. (Each process in itself could consist of
several subprocesses that have been distinguished in the parton-level
generator, but where this distinction is not carried along.)
\iteme{IDBMUP :}\label{p:IDBMUP} the PDG codes of the two incoming beam
particles (or, in alternative terminology, the beam and target particles).\\
In {\Py}, this replaces the information normally provided by the \ttt{BEAM}
and \ttt{TARGET} arguments of the \ttt{PYINIT} call. Only particles which
are acceptable \ttt{BEAM} or \ttt{TARGET} arguments may also be used in
\ttt{IDBMUP}. The {\galep} options are not available.
\iteme{EBMUP :}\label{p:EBMUP} the energies, in GeV, of the two incoming beam
particles. The first (second) particle is taken to travel in the $+z$ ($-z$)
direction.\\
The standard also allows non-collinear and varying-energy beams to be
specified, see \ttt{ISTUP = -9} below, but this is not yet implemented
in {\Py}.
\iteme{PDFGUP, PDFSUP :}\label{p:PDFGUP}\label{p:PDFSUP} the author group
(\ttt{PDFGUP}) and set (\ttt{PDFSUP}) of the parton distributions of the two
incoming beams, as used in the generation of the parton-level events.
Numbers are based on the \tsc{Pdflib} \cite{Plo93} lists, and should extend
to \tsc{LHAPDF} \cite{Gie02}. This enumeration
may not always be up to date, but it provides the only unique integer labels
for parton distributions that we have. Where no codes are yet assigned to the
parton distribution sets used, one should do as best as one can, and be
prepared for more extensive user interventions to interpret the information.
For lepton beams, or when the information is not provided for other
reasons, one should put \ttt{PDFGUP = PDFSUP = -1}.\\
By knowing which set has been used, it is possible to reweight cross sections
event by event, to correspond to another set.\\
Note that {\Py} does not access the \ttt{PDFGUP} or \ttt{PDFSUP} values
in its description of internal processes or initial-state showers. If you
want this to happen, you have to manipulate the \ttt{MSTP(51) - MSTP(56)}
switches. For instance, to access \tsc{Pdflib} for protons, put
\ttt{MSTP(51) = 1000*PDFGUP + PDFSUP} and \ttt{MSTP(52) = 2} in \ttt{UPINIT}.
(And remove the dummy \tsc{Pdflib} routines, as described for
\ttt{MSTP(52)}.) Also note that \ttt{PDFGUP} and \ttt{PDFSUP} allow an
independent choice of parton distributions on the two sides of the event,
whereas {\Py} only allows one single choice for all protons, another for
all pions and a third for all photons.\\
{\Py} implements one extension not specified in the LHA: If you set
\ttt{PDFGUP(i) = -9} for either of the two incoming beams, \ttt{i = 1}
or \ttt{= 2}, this signals that beam remnants are already included
in the specified final state and should not be provided by {\Py}. As a
consequence, neither initial-state radiation nor multiple interactions
are applied to the given particle configuration, since these would redefine
the beam remnants. What remains is resonance decays, final-state radiation,
hadronization and ordinary decays (unless explicitly switched off, of
course). One application could be to plug in parton-level configurations
already generated by some other initial-state shower algorithm. A more
typical example would be a generator for diffractive Higgs production,
$\p \p \to \p \p \H$, where {\Py} would be used to address the Higgs
decay with associated showers and handronization. Note that, in
accordance with the general rules, it is necessary to provide the
two incoming protons as the first two particles of the \ttt{/HEPEUP/}
event record, with status code $-1$. (Although this here happens to be
redundant, given the beam information provided at initialization.)
\iteme{IDWTUP :}\label{p:IDWTUP} master switch dictating how event weights
and cross sections should be interpreted. Several different models are
presented in detail below. There will be tradeoffs between these, e.g. a
larger flexibility to mix and re-mix several different processes could
require a larger administrative machinery. Therefore the best strategy
would vary, depending on the format of the input provided and the output
desired. In some cases, parton-level configurations have already been
generated with one specific model in mind, and then there may be no
choice.\\
\ttt{IDWTUP} significantly affects the interpretation of \ttt{XWGTUP},
\ttt{XMAXUP} and \ttt{XSECUP}, as described below, but the basic
nomenclature is the following. \ttt{XWGTUP} is the event weight for the
current parton-level event, stored in the
\ttt{HEPEUP} common block. For each allowed external process \ttt{i},
\ttt{XMAXUP(i)} gives the maximum event weight that could be encountered,
while \ttt{XSECUP(i)} is the cross section of the process. Here \ttt{i}
is an integer in the range between 1 and \ttt{NPRUP}; see the \ttt{LPRUP}
description below for comments on alternative process labels.
\begin{subentry}
\iteme{= 1 :} parton-level events come with a weight when input to {\Py},
but are then accepted or rejected, so that fully generated events at
output have a common weight,
customarily defined as $+1$. The event weight \ttt{XWGTUP} is a non-negative
dimensional quantity, in pb (converted to mb in {\Py}), with a mean value
converging to the total cross section of the respective process. For each
process \ttt{i}, the \ttt{XMAXUP(i)} value provides an upper estimate of how
large \ttt{XWGTUP} numbers can be encountered. There is no need to supply an
\ttt{XSECUP(i)} value; the cross sections printed with \ttt{PYSTAT(1)} are
based entirely on the averages of the \ttt{XWGTUP} numbers (with a small
correction for the fraction of events that \ttt{PYEVNT} fails to generate in
full for some reason).\\
The strategy is that \ttt{PYEVNT} selects which process \ttt{i} should be
generated next, based on the relative size of the \ttt{XMAXUP(i)} values.
The \ttt{UPEVNT} routine has to fill the \ttt{HEPEUP} common block with a
parton-level event of the requested type, and give its \ttt{XWGTUP} event
weight. The event is accepted by \ttt{PYEVNT} with probability
\ttt{XWGTUP/XMAXUP(i)}. In case of rejection, \ttt{PYEVNT} selects a new
process \ttt{i} and asks for a new event. This ensures that processes
are mixed in proportion to their average \ttt{XWGTUP} values.\\
This model presumes that \ttt{UPEVNT} is able to return a parton-level
event of the process type requested by \ttt{PYEVNT}. It works well if each
process is associated with an input stream of its own, either a subroutine
generating events `on the fly' or a file of already generated events. It
works less well if parton-level events from different processes already
are mixed in a
single file, and therefore cannot easily be returned in the order wanted by
\ttt{PYEVNT}. In the latter case one should either use another model or else
consider reducing the level of ambition: even if you have mixed several
different subprocesses on a file, maybe there is no need for {\Py} to
know this finer classification, in which case we may get back to a
situation with one `process' per external file. Thus the subdivision into
processes should be a matter of convenience, not a strait-jacket.
Specifically, the shower and hadronization treatment of a parton-level
event is independent of the process label assigned to it.
\\
If the events of some process are already available unweighted, then a
correct mixing of this process with others is ensured by putting
\ttt{XWGTUP = XMAXUP(i)}, where both of these numbers now is the total
cross section of the process.\\
Each \ttt{XMAXUP(i)} value must be known from the very beginning, e.g.
from an earlier exploratory run. If a larger value is encountered during
the course of the run, a warning message will be issued and the
\ttt{XMAXUP(i)} value (and its copy in \ttt{XSEC(ISUB,1)})
increased. Events generated before this time will
have been incorrectly distributed, both in the process composition and
in the phase space of the affected process, so that a bad estimate of
\ttt{XMAXUP(i)} may require a new run with a better starting value.\\
The model described here agrees with the one used for internal {\Py}
processes, and these can therefore freely be mixed with the external ones.
Internal processes are switched on with \ttt{MSUB(ISUB) = 1}, as usual,
either before the \ttt{PYINIT} call or in the \ttt{UPINIT} routine. One
cannot use \ttt{MSEL} to select a predefined set of processes, for
technical reasons, wherefore \ttt{MSEL = 0} is
hardcoded when external processes are included.\\
A reweighting of events is feasible, e.g. by including a
kinematics-dependent $K$ factor into \ttt{XWGTUP}, so long as
\ttt{XMAXUP(i)} is also properly modified to take this into account.
Optionally it is also possible to produce events with non-unit weight,
making use the \ttt{PYEVWT} facility, see section \ref{ss:PYwtevts}.
This works exactly the same way as for internal {\Py} processes,
except that the event information available inside \ttt{PYEVWT} would
be different for external processes. You may therefore wish to access
the \ttt{HEPEUP} common block inside your own copy of \ttt{PYEVWT},
where you calculate the event weight.\\
In summary, this option provides maximal flexibility, but at the price of
potentially requiring the administration of several separate input streams
of parton-level events.
\iteme{= -1 :} same as \ttt{= 1} above, except that event weights may be
either positive or negative on input, and therefore can come with an
output weight of $+1$ or $-1$. This weight is uniquely defined by the
sign of \ttt{XWGTUP}. It is also stored in \ttt{PARI(7)}.
The need for negative-weight events arises in some next-to-leading-order
calculations, but there are inherent dangers, discussed
in section \ref{sss:externcomm} below.\\
In order to allow a correct mixing between processes, a process of
indeterminate cross section sign has to be split up in two, where one
always gives a positive or vanishing \ttt{XWGTUP}, and the other always
gives it negative or vanishing. The \ttt{XMAXUP(i)} value for the latter
process should give the negative \ttt{XWGTUP} of largest magnitude that
will be encountered. \ttt{PYEVNT} selects which process \ttt{i} that
should be generated next, based on the relative size of the
\ttt{|XMAXUP(i)|} values. A given event is accepted with probability
\ttt{|XWGTUP|/|XMAXUP(i)|}.
\iteme{= 2 :} parton-level events come with a weight when input to {\Py},
but are then accepted or rejected, so that events at output have a common
weight, customarily defined as $+1$. The non-negative event weight
\ttt{XWGTUP} and its maximum value \ttt{XMAXUP(i)} may or may not be
dimensional quantities; it does not matter since only the ratio
\ttt{XWGTUP}$/$\ttt{XMAXUP(i)} will be used. Instead \ttt{XSECUP(i)} contains
the process cross section in pb (converted to mb in {\Py}). It is this
cross section that appears in the \ttt{PYSTAT(1)} table, only modified
by the small fraction of events that \ttt{PYEVNT} fails to generate in
full for some reason.\\
The strategy is that \ttt{PYEVNT} selects which process \ttt{i} should be
generated next, based on the relative size of the \ttt{XSECUP(i)} values.
The \ttt{UPEVNT} routine has to fill the \ttt{HEPEUP} common block with a
parton-level event of the requested type, and give its \ttt{XWGTUP} event
weight. The event is accepted by \ttt{PYEVNT} with probability
\ttt{XWGTUP/XMAXUP(i)}. In case of rejection, the process number \ttt{i}
is retained and \ttt{PYEVNT}
asks for a new event of this kind. This ensures that processes are mixed in
proportion to their \ttt{XSECUP(i)} values.\\
This model presumes that \ttt{UPEVNT} is able to return a parton-level
event of the process type requested by \ttt{PYEVNT}, with comments exactly
as for the \ttt{= 1} option.\\
If the events of some process are already available unweighted, then a
correct mixing of this process with others is ensured by putting
\ttt{XWGTUP = XMAXUP(i)}.\\
Each \ttt{XMAXUP(i)} and \ttt{XSECUP(i)} value must be known from the very
beginning, e.g.\ from an earlier integration run. If a larger value is
encountered during the course of the run, a warning message will be issued
and the \ttt{XMAXUP(i)} value
increased. This will not affect the process composition, but events
generated before this time will have been incorrectly distributed in the
phase space of the affected process, so that a bad estimate
of \ttt{XMAXUP(i)} may require a new run with a better starting value.\\
While the generation model is different from the normal internal {\Py} one,
it is sufficiently close that internal processes can be freely mixed
with the external ones, exactly as described for the \ttt{= 1} option. In
such a mix, internal processes are selected according to their equivalents
of \ttt{XMAXUP(i)} and at rejection a new \ttt{i} is selected, whereas
external ones are selected according to \ttt{XSECUP(i)} with \ttt{i}
retained when an event is rejected.\\
A reweighting of individual events is no longer simple, since this would
change the \ttt{XSECUP(i)} value nontrivially. Thus a new integration run
with the modified event weights would be necessary to obtain new
\ttt{XSECUP(i)} and \ttt{XMAXUP(i)} values. An overall rescaling of each
process separately can be obtained by modifying the \ttt{XSECUP(i)}
values accordingly, however, e.g.\ by a relevant $K$ factor.\\
In summary, this option is similar to the \ttt{= 1} one. The input of
\ttt{XSECUP(i)} allows good cross section knowledge also in short test runs,
but at the price of a reduced flexibility to reweight events.
\iteme{= -2 :} same as \ttt{= 2} above, except that event weights may be
either positive or negative on input, and therefore can come with an output
weight of $+1$ or $-1$. This weight is uniquely defined by the sign of
\ttt{XWGTUP}. It is also stored in \ttt{PARI(7)}.
The need for negative-weight events arises in some
next-to-leading-order calculations, but there are inherent dangers,
discussed in section \ref{sss:externcomm} below.\\
In order to allow a correct mixing between processes, a process of
indeterminate cross section sign has to be split up in two, where one
always gives a positive or vanishing \ttt{XWGTUP}, and the other always
gives it negative or vanishing. The \ttt{XMAXUP(i)} value for the latter
process should give the negative \ttt{XWGTUP} of largest magnitude that
will be encountered, and \ttt{XSECUP(i)} should give the
integrated negative cross section. \ttt{PYEVNT} selects which process
\ttt{i} that should be generated next, based on the relative size of the
\ttt{|XSECUP(i)|} values. A given event is accepted with probability
\ttt{|XWGTUP|/|XMAXUP(i)|}.
\iteme{= 3 :} parton-level events come with unit weight when input to {\Py},
\ttt{XWGTUP = 1}, and are thus always accepted. This makes the
\ttt{XMAXUP(i)} superfluous, while \ttt{XSECUP(i)} should give the
cross section of each process.\\
The strategy is that that the next process type \ttt{i} is selected by the
user inside \ttt{UPEVNT}, at the same time as the \ttt{HEPEUP} common block
is filled with information about the parton-level event. This event is then
unconditionally accepted by \ttt{PYEVNT}, except for the small fraction of
events that \ttt{PYEVNT} fails to generate in full for some reason.\\
This model allows \ttt{UPEVNT} to read events from a file where different
processes already appear mixed. Alternatively, you are free to devise and
implement your own mixing strategy inside \ttt{UPEVNT}, e.g. to mimic the
ones already outlined for \ttt{PYEVNT} in \ttt{= 1} and \ttt{= 2} above.\\
The \ttt{XSECUP(i)} values should be known from the beginning, in order for
\ttt{PYSTAT(1)} to produce a sensible cross section table. This is the only
place where it matters, however. That is, the processing of events inside
{\Py} is independent of this information.\\
In this model it is not possible to mix with internal {\Py} processes,
since not enough information is available to perform such a mixing.\\
A reweighting of events is completely in the hands of the \ttt{UPEVNT}
author. In the case that all events are stored in a single file, and all
are to be handed on to \ttt{PYEVNT}, only a common $K$ factor applied to all
processes would be possible.\\
In summary, this option puts more power --- and responsibility --- in the
hands of the author of the parton-level generator. It is very convenient for
the processing of unweighted parton-level events stored in a single file. The
price to be paid is a reduced flexibility in the reweighting of events,
or in combining processes at will.
\iteme{= -3 :} same as \ttt{= 3} above, except that event weights may be
either $+1$ or $-1$. This weight is uniquely defined by the sign of
\ttt{XWGTUP}. It is also stored in \ttt{PARI(7)}.
The need for negative-weight events arises in some
next-to-leading-order calculations, but there are inherent dangers,
discussed in section \ref{sss:externcomm} below.\\
Unlike the \ttt{= -1} and \ttt{= -2} options, there is no need to split a
process in two, each with a definite \ttt{XWGTUP} sign, since \ttt{PYEVNT}
is not responsible for the mixing of processes. It may well be that the
parton-level-generator author has enforced such a split, however, to solve a
corresponding mixing problem inside \ttt{UPEVNT}. Information on the relative
cross section in the negative- and positive-weight regions may also be useful
to understand the character and validity of the calculation (large
cancellations means trouble!).
\iteme{= 4 :} parton-level events come with a weight when input to {\Py},
and this weight is to be retained unchanged at output. The event weight
\ttt{XWGTUP} is a non-negative dimensional quantity, in pb (converted to
mb in {\Py}, and as such also stored in \ttt{PARI(7)}), with a mean value
converging to the total cross section of the respective process. When
histogramming results, one of these event weights would have to be used.\\
The strategy is exactly the same as \ttt{= 3} above, except that the event
weight is carried along from \ttt{UPEVNT} to the \ttt{PYEVNT} output. Thus
again all control is in the hands of the \ttt{UPEVNT} author.\\
A cross section can be calculated from the average of the \ttt{XWGTUP}
values, as in the \ttt{= 1} option, and is displayed by \ttt{PYSTAT(1)}.
Here it is of purely informative character, however, and does not influence
the generation procedure. Neither \ttt{XSECUP(i)} or \ttt{XMAXUP(i)} needs
to be known or supplied.\\
In this model it is not possible to mix with internal {\Py} processes,
since not enough information is available to perform such a mixing.\\
A reweighting of events is completely in the hands of the \ttt{UPEVNT}
author, and is always simple, also when events appear sequentially stored
in a single file.\\
In summary, this option allows maximum flexibility for the
parton-level-generator author, but potentially at the price of spending
a significant amount of time processing events of very small weight. Then
again, in some cases it may be an advantage to have more events in the
tails of a distribution in order to understand those tails better.
\iteme{= -4 :} same as \ttt{= 4} above, except that event weights in
\ttt{XWGTUP} may be either positive or negative. In particular, the mean
value of \ttt{XWGTUP} is converging to the total cross section of the
respective process. The need for negative-weight events arises in some
next-to-leading-order calculations, but there are inherent dangers,
discussed in section \ref{sss:externcomm} below.\\
Unlike the \ttt{= -1} and \ttt{= -2} options, there is no need to split
a process in two, each with a definite \ttt{XWGTUP} sign, since
\ttt{PYEVNT} does not have
to mix processes. However, as for option \ttt{= -3}, such a split may offer
advantages in understanding the character and validity of the calculation.
\end{subentry}
\iteme{NPRUP :}\label{p:NPRUP} the number of different external processes,
with information stored in the first \ttt{NPRUP} entries of the
\ttt{XSECUP}, \ttt{XERRUP}, \ttt{XMAXUP} and \ttt{LPRUP} arrays.
\iteme{XSECUP :}\label{p:XSECUP} cross section for each external process,
in pb. This information is mandatory for \ttt{IDWTUP =}~$\pm 2$, helpful
for $\pm 3$, and not used for the other options.
\iteme{XERRUP :}\label{p:XERRUP} the statistical error on the cross
section for each external process, in pb.\\
{\Py} will never make use of this information, but if it is available anyway
it provides a helpful service to the user of parton-level generators.\\
Note that, if a small number $n_{\mrm{acc}}$ of events pass the experimental
selection cuts, the statistical error on this cross section is limited by
$\delta \sigma / \sigma \approx 1/\sqrt{n_{\mrm{acc}}}$, irrespectively of
the quality of the original integration. Furthermore, at least in hadronic
physics, systematic errors from parton distributions and higher orders
typically are much larger than the statistical errors.
\iteme{XMAXUP :}\label{p:XMAXUP} the maximum event weight \ttt{XWGTUP} that
is likely to be encountered for each external process. For
\ttt{IDWTUP =}~$\pm 1$ it has dimensions pb, while the dimensionality need
not be specified for $\pm 2$. For the other \ttt{IDWTUP} options it is not
used.
\iteme{LPRUP :}\label{p:LPRUP} a unique integer identifier of each external
process, free to be picked by you for your convenience.
This code is used in the \ttt{IDPRUP} identifier of which process occured.\\
In {\Py}, an external process is thus identified by three different integers.
The first is the {\Py} process number, \ttt{ISUB}. This number is assigned
by \ttt{PYINIT} at the beginning of each run, by scanning the \ttt{ISET}
array for unused process numbers, and reclaiming such in the order they
are found. The second is the sequence number \ttt{i}, running from 1
through \ttt{NPRUP}, used to find information in the cross section arrays.
The third is the \ttt{LPRUP(i)} number, which can be anything that the user
wants to have as a unique identifier, e.g. in a larger database of processes.
For {\Py} to handle conversions, the two \ttt{KFPR} numbers of a given process
\ttt{ISUB} are overwritten with the second and third numbers above. Thus the
first external process will land in \ttt{ISUB = 4} (currently), and could
have \ttt{LPRUP(1) = 13579}. In a \ttt{PYSTAT(1)} call, it would be listed
as \ttt{User process 13579}.
\end{entry}
\subsubsection{Event information}
Inside the event loop of the main program, \ttt{PYEVNT} will be called to
generate the next event, as usual. When this is to be an external process,
the parton-level event configuration and the event weight is found by a call
from \ttt{PYEVNT} to \ttt{UPEVNT}.
\drawbox{CALL UPEVNT}\label{p:UPEVNT}
\begin{entry}
\itemc{Purpose:} routine to be provided by you when you want to implement
external processes, wherein the contents of the \ttt{HEPEUP} common block
are set. This information specifies the next parton-level event, and some
additional event information, see further below. How \ttt{UPEVNT} is expected
to solve its task depends on the model selected in \ttt{IDWTUP}, see above.
Specifically, note that the process type \ttt{IDPRUP} has already been
selected for some \ttt{IDWTUP} options (and then cannot be overwritten),
while it remains to be chosen for others.
\itemc{Note :} a dummy copy of \ttt{UPEVNT} is distributed with the program,
in order to avoid potential problems with unresolved external references.
This dummy should not be linked when you supply your own \ttt{UPEVNT}
routine. The code can be used to read event information previously
written by \ttt{PYUPEV}, however.
\end{entry}
\drawboxseven{~INTEGER MAXNUP}%
{~PARAMETER (MAXNUP=500)}%
{~INTEGER NUP,IDPRUP,IDUP,ISTUP,MOTHUP,ICOLUP}%
{~DOUBLE PRECISION XWGTUP,SCALUP,AQEDUP,AQCDUP,PUP,VTIMUP,SPINUP}%
{~COMMON/HEPEUP/NUP,IDPRUP,XWGTUP,SCALUP,AQEDUP,AQCDUP,IDUP(MAXNUP),}%
{\&ISTUP(MAXNUP),MOTHUP(2,MAXNUP),ICOLUP(2,MAXNUP),PUP(5,MAXNUP),}%
{\&VTIMUP(MAXNUP),SPINUP(MAXNUP)}
\begin{entry}\label{p:HEPEUP}
\itemc{Purpose :} to contain information on the latest external process
generated in \ttt{UPEVNT}. A part is one-of-a-kind numbers, like the event
weight, but the bulk of the information is a listing of incoming and
outgoing particles, with history, colour, momentum, lifetime and spin
information.
\iteme{MAXNUP :}\label{p:MAXNUP} the maximum number of particles that can
be specified by the external process.\\
The maximum of 500 is more than {\Py} is set up to handle. By default,
\ttt{MSTP(126) = 100}, at most 96 particles could be specified, since
4 additional entries are needed in {\Py} for the two beam particles
and the two initiators of initial-state radiation. If this default is
not sufficient, \ttt{MSTP(126)} would have to be increased at the
beginning of the run.
\iteme{NUP :}\label{p:NUP} the number of particle entries in the current
parton-level event, stored in the \ttt{NUP} first entries of the \ttt{IDUP},
\ttt{ISTUP}, \ttt{MOTHUP}, \ttt{ICOLUP}, \ttt{PUP}, \ttt{VTIMUP} and
\ttt{SPINUP} arrays. \\
The special value \ttt{NUP = 0} is used to denote the case where
\ttt{UPEVNT} is unable to provide an event, at least of the type requested
by \ttt{PYEVNT}, e.g. because all events available in a file have
already been read. For such an event also the error flag
\ttt{MSTI(51) = 1} instead of the normal \ttt{= 0}.
\iteme{IDPRUP :}\label{p:IDPRUP} the identity of the current process,
as given by the \ttt{LPRUP} codes.\\
When \ttt{IDWTUP =} $\pm 1$ or $\pm 2$, \ttt{IDPRUP} is selected by
\ttt{PYEVNT} and already set when entering \ttt{UPEVNT}. Then
\ttt{UPEVNT} has to provide an event of the specified process type,
but cannot change \ttt{IDPRUP}. When \ttt{IDWTUP =} $\pm 3$ or $\pm 4$,
\ttt{UPEVNT} is free to select the next process, and then should set
\ttt{IDPRUP} accordingly.
\iteme{XWGTUP :}\label{p:XWGTUP} the event weight. The precise definition
of \ttt{XWGTUP} depends on the value of the \ttt{IDWTUP} master switch.
For \ttt{IDWTUP = 1} or \ttt{= 4} it is a dimensional quantity, in pb, with
a mean value converging to the total cross section of the respective
process. For \ttt{IDWTUP = 2} the overall normalization is irrelevant.
For \ttt{IDWTUP = 3} only the value \ttt{+1} is allowed. For negative
\ttt{IDWTUP} also negative weights are allowed, although positive and
negative weights cannot appear mixed in the same process for
\ttt{IDWTUP = -1} or \ttt{= -2}.
\iteme{SCALUP :}\label{p:SCALUP} scale $Q$ of the event, as used in the
calculation of parton distributions (factorization scale). If the scale has
not been defined, this should be denoted by using the value \ttt{-1}.\\
In {\Py}, this is input to \ttt{PARI(21) - PARI(26)} (and internally
\ttt{VINT(51) - VINT(56)}) When \ttt{SCALUP} is non-positive, the
invariant mass of the parton-level event is instead used as scale.
Either of these comes to set the maximum virtuality in the initial-state
parton showers. The same scale is also used for the first final-state
shower, i.e.\ the one associated with the hard scattering. As in
internal events, \ttt{PARP(67)} and \ttt{PARP(71)} offer multiplicative
factors, whereby the respective initial- or final-state showering
$Q_{\mrm{max}}^2$ scale can be modified relative to the scale above.
Any subsequent final-state showers are assumed to come from resonance
decays, where the resonance mass always sets the scale. Since \ttt{SCALUP}
is not directly used inside {\Py} to evaluate parton densities, its role
as regulator of parton-shower activity may be the more important one.
\iteme{AQEDUP :}\label{p:AQEDUP} the QED coupling $\alphaem$ used for this
event. If $\alphaem$ has not been defined, this should be denoted by using
the value \ttt{-1}.\\
In {\Py}, this value is stored in \ttt{VINT(57)}. It is not used anywhere,
however.
\iteme{AQCDUP :}\label{p:AQCDUP} the QCD coupling $\alphas$ used for this
event. If $\alphas$ has not been defined, this should be denoted by using
the value \ttt{-1}.\\
In {\Py}, this value is stored in \ttt{VINT(58)}. It is not used anywhere,
however.
\iteme{IDUP(i) :}\label{p:IDUP} particle identity code, according to the
PDG convention, for particle \ttt{i}. As an extension to this standard,
\ttt{IDUP(i) = 0} can be used to designate an intermediate state of
undefined (and possible non-physical) character, e.g.\ a subsystem with
a mass to be preserved by parton showers.\\
In the {\Py} event record, this corresponds to the \ttt{KF = K(I,2)} code.
But note that, here and in the following, the positions \ttt{i} in
\ttt{HEPEUP} and \ttt{I} in \ttt{PYJETS}
are likely to be different, since {\Py} normally stores more information in
the beginning of the event record. Since \ttt{K(I,2) = 0} is forbidden, the
\ttt{IDUP(i) = 0} code is mapped to \ttt{K(I,2) = 90}.
\iteme{ISTUP(i) :}\label{p:ISTUP} status code of particle \ttt{i}.
\begin{subentry}
\iteme{= -1 :} an incoming particle of the hard-scattering process.\\
In {\Py}, currently it is presumed that the first two particles,
\ttt{i = 1} and \ttt{= 2}, are of this character, and none of the others.
If this is not the case, the \ttt{HEPEUP} record will be rearranged to
put such entries first. If the listing is still not acceptable after this,
the program execution will stop.
This is a restriction relative to the standard, which allows more
possibilities. It is also presumed that these two particles are given with
vanishing masses and parallel to the respective incoming beam direction,
i.e.\ $E = p_z$ for the first and $E = - p_z$ for the second. Should the
particles not be massless at input, $E$ and $p_z$ is shuffled between the
two incoming partons to assure this, while preserving the total quantities.
The assignment of space-like virtualities and nonvanishing $\pT$'s from
initial-state radiation and primordial $\kT$'s is the prerogative of {\Py}.
\iteme{= 1 :} an outgoing final-state particle.\\
Such a particle can, of course, be processed further by {\Py}, to add
showers and hadronization, or perform decays of any remaining resonances.
\iteme{= 2 :} an intermediate resonance, whose mass should be preserved by
parton showers. For instance, in a process such as
$\e^+\e^- \to \Z^0 \hrm^0 \to \q\qbar \, \b\bbar$, the $\Z^0$ and $\hrm^0$
should both be flagged this way, to denote that the $\q\qbar$ and
$\b\bbar$ systems should have their individual masses preserved.
In a more complex example, $\d \ubar \to \W^-\Z^0\g \to %
\ell^-\br{\nu}_{\ell} \, \q \qbar \, \g$, both the $\W^-$ and
$\Z^0$ particles and the $\W^-\Z^0$ pseudoparticle (with \ttt{IDUP(i) = 0})
could be given with status 2.\\
Often mass preservation is correlated with colour singlet subsystems,
but this need not be the case. In
$\e^+ \e^- \to \t \tbar \to \b \W^+ \, \bbar \W^-$, the $\b$ and
$\bbar$ would be in a colour singlet state, but \textit{not} with a
preserved mass. Instead the $\t = \b \W^+$ and $\tbar = \bbar \W^-$
masses would be preserved, i.e.\ when $\b$ radiates $\b \to \b \g$
the recoil is taken by the $\W^+$. Exact mass preservation also by the
hadronization stage is only guaranteed for colour singlet subsystems,
however, at least for string fragmentation, since it is not possible to
define a subset of hadrons that uniquely belong only with a single
coloured particle.\\
The assignment of intermediate states is not always quantum mechanically
well-defined. For instance,
$\e^+\e^- \to \mu^- \mu^+ \nu_{\mu} \br{\nu}_{\mu}$ can proceed
both through a $\W^+\W^-$ and a $\Z^0\Z^0$ intermediate state, as well
as through other graphs, which can interfere with each other. It is here
the responsibility of the matrix-element-generator author to pick one
of the alternatives, according to some convenient recipe. One option
might be to perform two calculations, one complete to select the event
kinematics and calculate the event weight, and a second with all
interference terms neglected to pick the event history according to the
relative weight of each graph. Often one particular graph would dominate,
because a certain pairing of the final-state fermions would give invariant
masses on or close to some resonance peaks.\\
In {\Py}, the identification of an intermediate resonance is not only a
matter of preserving a mass, but also of improving the modelling of the
final-state shower evolution, since matrix-element-correction factors
have been calculated for a variety of possible resonance decays and
implemented in the respective parton-shower description, see section
\ref{sss:PSMEfsmerge}.
\iteme{= 3 :} an intermediate resonance, given for documentation only,
without any demand that the mass should be preserved in subsequent
showers.\\
In {\Py}, currently particles defined with this option are not treated
any differently from the ones with \ttt{= 2}.
\iteme{= -2 :} an intermediate space-like propagator, defining an $x$
and a $Q^2$, in the Deeply Inelastic Scattering terminology, which
should be preserved.\\
In {\Py}, currently this option is not defined and should not be used.
If it is, the program execution will stop.
\iteme{= -9 :} an incoming beam particle at time $t = -\infty$. Such
beams are not required in most cases, since the \ttt{HEPRUP} common
block normally contains the information. The exceptions are
studies with non-collinear beams and with varying-energy beams
(e.g.\ from beamstrahlung, section \ref{sss:estructfun}), where
\ttt{HEPRUP} does not supply sufficient flexibility. Information given
with \ttt{= -9} overwrites the one given in \ttt{HEPRUP}.\\
This is an optional part of the standard, since it may be difficult
to combine with some of the \ttt{IDWTUP} options.\\
Currently it is not recognized by {\Py}. If it is used, the program
execution will stop.
\end{subentry}
\iteme{MOTHUP(1,i), MOTHUP(2,i) :}\label{p:MOTHUP} position of the
first and last mother of particle \ttt{i}. Decay products will
normally have only one mother. Then either \ttt{MOTHUP(2,i) = 0} or
\ttt{MOTHUP(2,i) = MOTHUP(1,i)}. Particles in the outgoing state of a
$2 \to n$ process have two mothers. This scheme does not limit the
number of mothers, so long as these appear consecutively in the listing,
but in practice there will likely never be more than two mothers per
particle.\\
As has already been mentioned for \ttt{ISTUP(i) = 2}, the definition of
history is not always unique. Thus, in a case like
$\e^+\e^- \to \mu^+ \mu^- \gamma$, proceeding via an intermediate
$\gamma^*/\Z^0$, the squared matrix element contains an interference
term between initial- and final-state emission of the photon. This
ambiguity has to be resolved by the matrix-elements-based generator.\\
In {\Py}, only information on the first mother survives into
\ttt{K(I,3)}. This is adequate for resonance decays, while particles
produced in the primary $2 \to n$ process are given mother code 0,
as is customary for internal processes. It implies that two particles
are deemed to have the same mothers if the first one agrees; it is
difficult to conceive of situations where this would not be the case.
Furthermore, it is assumed that the \ttt{MOTHUP(1,i) < i}, i.e.\ that
mothers are stored ahead of their daughters, and that all daughters of
a mother are listed consecutively, i.e.\ without other particles
interspersed. If this is not the case, the \ttt{HEPEUP} record will
be rearranged so as to adhere to these principles. \\
{\Py} has a limit of at most 80 particles coming from the same mother,
for the final-state parton-shower algorithm to work. In fact, the shower
is optimized for a primary $2 \to 2$ process followed by some sequence
of $1 \to 2$ resonance decays. Then colour coherence with the initial
state, matrix-element matching to gluon emission in resonance decays,
and other sophisticated features are automatically included. By contrast,
the description of emission in systems with three or more partons is
less sophisticated. Apart from problems with the algorithm itself,
more information than is provided by the standard would be needed to
do a good job. Specifically, there is a significant danger of
double-counting or gaps between the radiation already covered by matrix
elements and the one added by the shower. The omission from \ttt{HEPEUP}
of intermediate resonances known to be there, so that e.g. two
consecutive $1 \to 2$ decays are bookkept as a single $1 \to 3$
branching, is a simple way to reduce the reliability of your studies!
\iteme{ICOLUP(1,i), ICOLUP(2,i) :}\label{p:ICOLUP} integer tags for the
colour flow lines passing through the colour and anticolour, respectively,
of the particle. Any particle with colour (anticolour), such as a quark
(antiquark) or gluon, will have the first (second) number nonvanishing.\\
The tags can be viewed as a numbering of different colours in the
$N_C \to \infty$ limit of QCD. Any nonzero integer can be used to
represent a different colour, but the standard recommends to stay
with positive numbers larger than \ttt{MAXNUP} to avoid confusion
between colour tags and the position labels \ttt{i} of particles.\\
The colour and anticolour of a particle is defined with the respect
to the physical time ordering of the process, so as to allow a unique
definition of colour flow also through intermediate particles. That is,
a quark always has a nonvanishing colour tag \ttt{ICOLUP(1,i)}, whether
it is in the initial, intermediate or final state. A simple example would
be $\q\qbar \to \t\tbar \to \b \W^+ \bbar \W^-$, where the same colour
label is to be used for the $\q$, the $\t$ and the $\b$. Correspondingly,
the $\qbar$, $\tbar$ and $\bbar$ share another colour label, now stored
in the anticolour position \ttt{ICOLUP(2,i)}.\\
The colour label in itself does not distinguish between the colour or
the anticolour of a given kind; that information appears in the usage
either of the \ttt{ICOLUP(1,i)} or of the \ttt{ICOLUP(2,i)} position
for the colour or anticolour, respectively. Thus, in a $\W^+ \to \u \dbar$
decay, the $\u$ and $\dbar$ would share the same colour label, but stored
in \ttt{ICOLUP(1,i)} for the $\u$ and in \ttt{ICOLUP(2,i)} for the
$\dbar$.\\
In general, several colour flows are possible in a given subprocess.
This leads to ambiguities, of a character similar to the ones for the
history above, and as is discussed in section \ref{sss:QCDjetclass}.
Again it is up to the author of the matrix-elements-based generator to
find a sensible solution. It is useful to note that all interference terms
between different colour flow topologies vanish in the $N_C \to \infty$
limit of QCD. One solution would then be to use a first calculation in
standard QCD to select the momenta and find the weight of the process,
and a second with $N_C \to \infty$ to pick one specific colour topology
according to the relative probabilities in this limit.\\
The above colour scheme also allows for baryon-number-violating processes.
Such a vertex would appear as `dangling' colour lines, when the
\ttt{ICOLUP} and \ttt{MOTHUP} information is correlated.
For instance, in $\su \to \dbar \dbar$ the $\su$ inherits an existing
colour label, while the two $\dbar$'s are produced with two different
new labels.\\
Several examples of colour assignments, both with and without baryon
number violation, are given in \cite{Boo01}.\\
In {\Py}, baryon number violation is not yet implemented as part of the
external-process machinery (but exists for internal processes, including
internally handled decays of resonances provided externally). It will
require substantial extra work to lift this restriction, and this is not
imminent.
\iteme{PUP(1,i), PUP(2,i), PUP(3,i), PUP(4,i), PUP(5,i) :}\label{p:PUP}
the particle momentum vector $(p_x, p_y, p_z, E, m)$, with units of GeV.
A space-like virtuality is denoted by a negative sign on the mass.\\
Apart from the index order, this exactly matches the \ttt{P} momentum
conventions in \ttt{PYJETS}.\\
{\Py} is forgiving when it comes to using other masses than
its own, e.g.\ for quarks. Thus the external process can be given
directly with the $m_{\b}$ used in the calculation, without any worry
how this matches the {\Py} default. However, remember that the two
incoming particles with \ttt{ISTUP(i) = -1} have to be massless,
and will be modified correspondingly if this is not the case at input.
\iteme{VTIMUP(i) :}\label{p:VTIMUP} invariant lifetime $c\tau$ in mm,
i.e.\ distance from production to decay. Once the primary vertex has
been selected, the subsequent decay vertex positions in space and time
can be constructed step by step, by also making use of the momentum
information. Propagation in vacuum, without any bending e.g.\ by
magnetic fields, has to be assumed.\\
This exactly corresponds to the \ttt{V(I,5)} component in \ttt{PYJETS}.
Note that it is used in {\Py} to track colour singlet particles through
distances that might be observable in a detector. It is not used to
trace the motion of coloured partons at fm scales, through the
hadronization process. Also note that {\Py} will only use this
information for intermediate resonances, not for the
initial- and final-state particles. For instance, for an undecayed
$\tau^-$, the lifetime is selected as part of the $\tau^-$ decay
process, not based on the \ttt{VTIMUP(i)} value.
\iteme{SPINUP(i) :}\label{p:SPINUP} cosine of the angle between the
spin vector of a particle and its three-momentum, specified in the
lab frame, i.e.\ the frame where the event as a whole is defined.
This scheme is neither general nor complete, but it is chosen as a
sensible compromise.\\
The main foreseen application is $\tau$'s with a specific helicity.
Typically a relativistic $\tau^-$ ($\tau^+$) coming from a $\W^-$
($\W^+$) decay would have helicity and \ttt{SPINUP(i) =} $-1$ ($+1$).
This could be changed by the boost from the $\W$ rest frame to the
lab frame, however. The use of a real number, rather than an integer,
allows for an extension to the non-relativistic case.\\
Particles which are unpolarized or have unknown polarization should be
given \ttt{SPINUP(i) = 9}.\\
Explicit spin information is not used anywhere in {\Py}. It
is implicit in many production and decay matrix elements, which often
contain more correlation information than could be conveyed by the
simple spin numbers discussed here. Correspondingly, it is to be
expected that the external generator already performed the decays of
the $\W$'s, the $\Z$'s and the other resonances, so as to include the
full spin correlations. If this is not the case, such resonances will
normally be decayed isotropically. Some correlations could appear in
decay chains: the {\Py} decay $\t \to \b\W^+$ is isotropic, but the
subsequent $\W^+ \to \q_1 \qbar_2$ decay contains implicit $\W$ helicity
information from the $\t$ decay.\\
Also $\tau$ decays performed by {\Py} would be isotropic. An interface
routine \ttt{PYTAUD} (see section \ref{ss:JETphysrout}) can be used
to link to external $\tau$ decay generators, but is based on defining
the $\tau$ in the rest frame of the decay that produces it, and so is
not directly applicable here. It could be rewritten to make use of the
\ttt{SPINUP(i)} information, however. In the meantime, and of
course also afterwards, a valid option is to perform the $\tau$ decays
yourself before passing `parton-level' events to {\Py}.
\end{entry}
One auxiliary routine exists, that formally is part of the
{\Py} package, but could be used by any generator:
\begin{entry}
\iteme{SUBROUTINE PYUPRE :}\label{p:PYUPRE}
called immediately after \ttt{UPEVNT} has been called to provide a
user-process event. It will rearrange the contents of the
\ttt{HEPEUP} common block so that afterwards the two incoming
partons appear in lines 1 and 2, so that all mothers appear ahead
of their daughters, and so that the daughters of a decay are listed
consecutively. Such an order can thereby be presumed to exist in
the subsequent parsing of the event. If the rules already are
obeyed, the routine does not change the order. Further, the routine
can shuffle energy and momentum between the two incoming partons
to ensure that they are both massless.
\end{entry}
\subsubsection{An example}
To exemplify the above discussion, consider the explicit case of
$\q\qbar$ or $\g\g \to \t\tbar \to \b\W^+\bbar\W^- \to %
\b\q_1\qbar_2 \, \bbar\q_3\qbar_4$. These two processes are already
available in {\Py}, but without full spin correlations. One might
therefore wish to include them from some external generator. A physics
analysis would then most likely involve angular correlations intended
to set limits on (or reveal traces of) anomalous couplings. However, so
as to give a simple self-contained example, instead consider the analysis
of the charged multiplicity distribution. This actually offers a simple
first cross-check between the internal and external implementations of
the same process. The main program might then look something like
\begin{verbatim}
IMPLICIT DOUBLE PRECISION(A-H, O-Z)
IMPLICIT INTEGER(I-N)
C...User process event common block.
INTEGER MAXNUP
PARAMETER (MAXNUP=500)
INTEGER NUP,IDPRUP,IDUP,ISTUP,MOTHUP,ICOLUP
DOUBLE PRECISION XWGTUP,SCALUP,AQEDUP,AQCDUP,PUP,VTIMUP,SPINUP
COMMON/HEPEUP/NUP,IDPRUP,XWGTUP,SCALUP,AQEDUP,AQCDUP,IDUP(MAXNUP),
&ISTUP(MAXNUP),MOTHUP(2,MAXNUP),ICOLUP(2,MAXNUP),PUP(5,MAXNUP),
&VTIMUP(MAXNUP),SPINUP(MAXNUP)
SAVE /HEPEUP/
C...PYTHIA common block.
COMMON/PYJETS/N,NPAD,K(4000,5),P(4000,5),V(4000,5)
SAVE /PYJETS/
C...Initialize.
CALL PYINIT('USER',' ',' ',0D0)
C...Book histogram. Reset event counter.
CALL PYBOOK(1,'Charged multiplicity',100,-1D0,199D0)
NACC=0
C...Event loop; check that not at end of run; list first events.
DO 100 IEV=1,1000
CALL PYEVNT
IF(NUP.EQ.0) GOTO 110
NACC=NACC+1
IF(IEV.LE.3) CALL PYLIST(7)
IF(IEV.LE.3) CALL PYLIST(2)
C...Analyse event; end event loop.
CALL PYEDIT(3)
CALL PYFILL(1,DBLE(N),1D0)
100 CONTINUE
C...Statistics and histograms.
110 CALL PYSTAT(1)
CALL PYFACT(1,1D0/DBLE(NACC))
CALL PYHIST
END
\end{verbatim}
There \ttt{PYINIT} is called with \ttt{'USER'} as first argument,
implying that the rest is dummy. The event loop itself looks fairly
familiar, but with two additions. One is that \ttt{NUP} is checked
after each event, since \ttt{NUP = 0} would signal a premature end
of the run, with the external generator unable to return more events.
This would be the case e.g.\ if events have been stored on file, and
the end of this file is reached. The other is that \ttt{CALL PYLIST(7)}
can be used to list the particle content of the \ttt{HEPEUP} common
block (with some information omitted, on vertices and spin), so that
one can easily compare this input with the output after {\Py}
processing, \ttt{CALL PYLIST(2)}. An example of a \ttt{PYLIST(7)}
listing would be
\begin{verbatim}
Event listing of user process at input (simplified)
I IST ID Mothers Colours p_x p_y p_z E m
1 -1 21 0 0 101 109 0.000 0.000 269.223 269.223 0.000
2 -1 21 0 0 109 102 0.000 0.000 -225.566 225.566 0.000
3 2 6 1 2 101 0 72.569 153.924 -10.554 244.347 175.030
4 2 -6 1 2 0 102 -72.569 -153.924 54.211 250.441 175.565
5 1 5 3 0 101 0 56.519 33.343 53.910 85.045 4.500
6 2 24 3 0 0 0 16.050 120.581 -64.464 159.302 80.150
7 1 -5 4 0 0 102 44.127 -60.882 25.507 79.527 4.500
8 2 -24 4 0 0 0 -116.696 -93.042 28.705 170.914 78.184
9 1 2 6 0 103 0 -8.667 11.859 16.063 21.766 0.000
10 1 -1 6 0 0 103 24.717 108.722 -80.527 137.536 0.000
11 1 -2 8 0 0 104 -33.709 -22.471 -26.877 48.617 0.000
12 1 1 8 0 104 0 -82.988 -70.571 55.582 122.297 0.000
\end{verbatim}
Note the reverse listing of \ttt{ID(UP)} and \ttt{IST(UP)} relative to the
\ttt{HEPEUP} order, to have better agreement with the \ttt{PYJETS} one.
(The \ttt{ID} column is wider in real life, to allow for longer codes,
but has here been reduced to fit the listing onto the page.)
The corresponding \ttt{PYLIST(2)} listing of course would be considerably
longer, containing a complete event as it does. Also the particles above
would there appear boosted by the effects of initial-state radiation and
primordial $\kT$; copies of them further down in the event record would
also include the effects of final-state radiation. The full story is
available with \ttt{MSTP(125) = 2}, while the default listing omits some
of the intermediate steps.
The \ttt{PYINIT} call will generate a call to the user-supplied
routine \ttt{UPINIT}. It is here that we need to specify the details
of the generation model. Assume, for instance, that $\q\qbar$- and
$\g\g$-initiated events have been generated in two separate runs
for Tevatron Run II, with weighted events stored in two separate files.
By the end of each run, cross section and maximum weight information
has also been obtained, and stored on separate files. Then
\ttt{UPINIT} could look like
\begin{verbatim}
SUBROUTINE UPINIT
C...Double precision and integer declarations.
IMPLICIT DOUBLE PRECISION(A-H, O-Z)
IMPLICIT INTEGER(I-N)
C...User process initialization common block.
INTEGER MAXPUP
PARAMETER (MAXPUP=100)
INTEGER IDBMUP,PDFGUP,PDFSUP,IDWTUP,NPRUP,LPRUP
DOUBLE PRECISION EBMUP,XSECUP,XERRUP,XMAXUP
COMMON/HEPRUP/IDBMUP(2),EBMUP(2),PDFGUP(2),PDFSUP(2),
&IDWTUP,NPRUP,XSECUP(MAXPUP),XERRUP(MAXPUP),XMAXUP(MAXPUP),
&LPRUP(MAXPUP)
SAVE /HEPRUP/
C....Pythia common block - needed for setting PDF's; see below.
COMMON/PYPARS/MSTP(200),PARP(200),MSTI(200),PARI(200)
SAVE /PYPARS/
C...Set incoming beams: Tevatron Run II.
IDBMUP(1)=2212
IDBMUP(2)=-2212
EBMUP(1)=1000D0
EBMUP(2)=1000D0
C...Set PDF's of incoming beams: CTEQ 5L.
C...Note that Pythia will not look at PDFGUP and PDFSUP.
PDFGUP(1)=4
PDFSUP(1)=46
PDFGUP(2)=PDFGUP(1)
PDFSUP(2)=PDFSUP(1)
C...Set use of CTEQ 5L in internal Pythia code.
MSTP(51)=7
C...If you want Pythia to use PDFLIB, you have to set it by hand.
C...(You also have to ensure that the dummy routines
C...PDFSET, STRUCTM and STRUCTP in Pythia are not linked.)
C MSTP(52)=2
C MSTP(51)=1000*PDFGUP(1)+PDFSUP(1)
C...Decide on weighting strategy: weighted on input, cross section known.
IDWTUP=2
C...Number of external processes.
NPRUP=2
C...Set up q qbar -> t tbar.
OPEN(21,FILE='qqtt.info',FORM='unformatted',ERR=100)
READ(21,ERR=100) XSECUP(1),XERRUP(1),XMAXUP(1)
LPRUP(1)=661
OPEN(22,FILE='qqtt.events',FORM='unformatted',ERR=100)
C...Set up g g -> t tbar.
OPEN(23,FILE='ggtt.info',FORM='unformatted',ERR=100)
READ(23,ERR=100) XSECUP(2),XERRUP(2),XMAXUP(2)
LPRUP(2)=662
OPEN(24,FILE='ggtt.events',FORM='unformatted',ERR=100)
RETURN
C...Stop run if file operations fail.
100 WRITE(*,*) 'Error! File open or read failed. Program stopped.'
STOP
END
\end{verbatim}
Here unformatted read/write is used to reduce the size of the
event files, but at the price of a platform dependence. Formatted files
are preferred if they are to be shipped elsewhere. The rest should be
self-explanatory.
Inside the event loop of the main program, \ttt{PYEVNT} will call
\ttt{UPEVNT} to obtain the next parton-level event. In its simplest
form, only a single \ttt{READ} statement would be necessary to read
information on the next event, e.g.\ what is shown in the event
listing earlier in this section, with a few additions. Then the
routine could look like
\begin{verbatim}
SUBROUTINE UPEVNT
C...Double precision and integer declarations.
IMPLICIT DOUBLE PRECISION(A-H, O-Z)
IMPLICIT INTEGER(I-N)
C...User process event common block.
INTEGER MAXNUP
PARAMETER (MAXNUP=500)
INTEGER NUP,IDPRUP,IDUP,ISTUP,MOTHUP,ICOLUP
DOUBLE PRECISION XWGTUP,SCALUP,AQEDUP,AQCDUP,PUP,VTIMUP,SPINUP
COMMON/HEPEUP/NUP,IDPRUP,XWGTUP,SCALUP,AQEDUP,AQCDUP,IDUP(MAXNUP),
&ISTUP(MAXNUP),MOTHUP(2,MAXNUP),ICOLUP(2,MAXNUP),PUP(5,MAXNUP),
&VTIMUP(MAXNUP),SPINUP(MAXNUP)
SAVE /HEPEUP/
C...Pick file to read from, based on requested event type.
IUNIT=22
IF(IDPRUP.EQ.662) IUNIT=24
C...Read event from this file. (Except that NUP and IDPRUP are known.)
NUP=12
READ(IUNIT,ERR=100,END=100) XWGTUP,SCALUP,AQEDUP,AQCDUP,
&(IDUP(I),ISTUP(I),MOTHUP(1,I),MOTHUP(2,I),ICOLUP(1,I),
&ICOLUP(2,I),(PUP(J,I),J=1,5),VTIMUP(I),SPINUP(I),I=1,NUP)
C...Return, with NUP=0 if read failed.
RETURN
100 NUP=0
RETURN
END
\end{verbatim}
However, in reality one might wish to save disk space by not storing
redundant information. The \ttt{XWGTUP} and \ttt{SCALUP} numbers are
vital, while \ttt{AQEDUP} and \ttt{AQCDUP} are purely informational
and can be omitted. In a $\g\g \to \t\tbar \to \b\W^+\bbar\W^- \to %
\b\q_1\qbar_2 \, \bbar\q_3\qbar_4$ event, only the $\q_1$, $\qbar_2$,
$\q_3$ and $\qbar_4$ flavours need be given, assuming that the particles
are always stored in the same order. For a $\q\qbar$ initial state,
the $\q$ flavour should be added to the list. The \ttt{ISTUP},
\ttt{MOTHUP} and \ttt{ICOLUP} information is the same in all events
of a given process, except for a twofold ambiguity in the colour flow
for $\g\g$ initial states. All \ttt{VTIMUP} vanish and
the \ttt{SPINUP} are uninteresting since spins have already been taken
into account by the kinematics of the final fermions. (It would be
different if one of the $\W$'s decayed leptonically to a $\tau$.)
Only the \ttt{PUP} values of the six final fermions need be given,
since the other momenta and masses can be reconstructed from this,
remembering that the two initial partons are massless and along the
beam pipe. The final fermions are on the mass shell, so their masses
need not be stored event by event, but can be read from a small table.
The energy of a particle can be
reconstructed from its momentum and mass. Overall transverse momentum
conservation removes two further numbers. What remains is thus 5 integers
and 18 real numbers, where the reals could well be stored in single
precision. Of course, the code needed to unpack information stored
this way would be lengthy but trivial. Even more compact storage
strategies could be envisaged, e.g. only to save the weight and the
seed of a dedicated random-number generator, to be used to generate
the next parton-level event. It is up to you to find
the optimal balance between disk space and coding effort.
\subsubsection{PYTHIA as a generator of external processes}
It is possible to write {\Py}-generated hard processes to disk,
and then read them back in for the generation of a complete event,
making use of the LHA conventions. This facility may be
of limited practical usefulness, and is mainly intended for debug
purposes. Thus there are some limitations to what it can do,
especially no pileup and no photon beams (including the
\ttt{'gamma/l'} options), and no weighted events. Basically, it is
intended for hard processes in $\p\p / \p\pbar / \e^+\e^-$ collisions.
Events are stored using the \ttt{IDWTUP = 3} strategy, i.e.\ already
mixed and reweighted, so no rejection step is needed.
To generate the appropriate files, you have to:
\begin{Enumerate}
\item Open one file where the initial run information in \ttt{HEPRUP}
can be stored, and give its file number in \ttt{MSTP(161)}.
\item Open one file where events in the \ttt{HEPEUP} format can be
stored, and give its file number in \ttt{MSTP(162)}.
\item Define hard processes as normally, and initialize with \ttt{PYINIT}.
\item In the event loop, replace the call to \ttt{PYEVNT}
(or \ttt{PYEVNW}) by a call to \ttt{PYUPEV}\label{p:PYUPEV}, which
will generate the hard process as usual, and write this process onto
file number \ttt{MSTP(162)}.
\item After all events have been generated, call
\ttt{PYUPIN}\label{p:PYUPIN} to store the run information. (The name of
this routine may seem inappropriate, since it is not called at the
initialization but at the end. However, the name refers to this
information being read in first in the next step, see below).
\end{Enumerate}
Once you have these files, you can use them in a later run to
generate complete events. This run follows the normal pattern for
user-defined processes, with only minor additions.
\begin{Enumerate}
\item Open the file where the initial run information in \ttt{HEPRUP}
was stored, and give its file number in \ttt{MSTP(161)}.
\item Open the file where events in the \ttt{HEPEUP} format were
stored, and give its file number in \ttt{MSTP(162)}.
\item Initialize as usual with \ttt{CALL PYINIT('USER',' ',' ',0D0)}.
The 'dummy' \ttt{UPINIT} routine in the Pythia library contains the
code for reading the initial run information written by \ttt{PYUPIN}
(but not in any other format), so in this case you need not supply any
routine of your own.
\item In the event loop, call \ttt{PYEVNT} (or \ttt{PYEVNW}) as usual
to generate the next event. The 'dummy' \ttt{UPEVNT} routine in the
{\Py} library contains the code for reading the event information
written by \ttt{PYUPEV} (but not in any other format), so in this case
you need not supply any routine of your own.
\end{Enumerate}
\subsubsection{Further comments}
\label{sss:externcomm}
This section contains additional information on a few different
topics: cross section interpretation, negative-weight events,
relations with other {\Py} switches and routines,
and error conditions.
In several \ttt{IDWTUP} options,
the \ttt{XWGTUP} variable is supposed to give the differential
cross section of the current event, times the phase-space volume
within which events are generated, expressed in picobarns.
(Converted to millibarns inside {\Py}.) This means that, in the limit
that many events are generated, the average value of \ttt{XWGTUP}
gives the total cross section of the simulated process.
Of course, the tricky part is that the differential cross section
usually is strongly peaked in a few regions of the phase space, such
that the average probability to accept an event,
$\langle$\ttt{XWGTUP}$\rangle /$\ttt{XMAXUP(i)} is
small. It may then be necessary to find a suitable set of transformed
phase-space coordinates, for which the correspondingly transformed
differential cross section is better behaved.
To avoid confusion, here is a more formal version of the above
paragraph. Call $\d X$ the differential phase space, e.g.\ for a
$2 \to 2$
process $\d X = \d x_1 \, \d x_2 \, \d \hat{t}$, where $x_1$ and $x_2$
are the momentum fractions carried by the two incoming partons and
$\hat{t}$ the Mandelstam variable of the scattering (see section
\ref{ss:kinemtwo}). Call $\d \sigma / \d X$ the differential cross section
of the process, e.g.\ for $2 \to 2$: $\d \sigma / \d X = \sum_{ij}
f_i(x_1,Q^2) \, f_j(x_2,Q^2) \, \d \hat{\sigma}_{ij} / \d \hat{t}$,
i.e.\ the product of parton distributions and hard-scattering matrix
elements, summed over all allowed incoming flavours $i$ and $j$.
The physical cross section that one then wants to generate is
$\sigma = \int (\d \sigma / \d X) \, \d X$, where the integral is over
the allowed phase-space volume. The event generation procedure consists
of selecting an $X$ uniformly in $\d X$ and then evaluating the weight
$\d \sigma / \d X$ at this point. \ttt{XWGTUP} is now simply
\ttt{XWGTUP}$ = \d \sigma / \d X \, \int \d X$, i.e.\ the differential
cross section times the considered volume of phase space. Clearly,
when averaged over many events, \ttt{XWGTUP} will correctly estimate
the desired cross section. If \ttt{XWGTUP} fluctuates too much, one
may try to transform to new variables $X'$,
where events are now picked accordingly to $\d X'$ and
\ttt{XWGTUP}$ = \d \sigma / \d X' \, \int \d X'$.
A warning. It is important that $X$ is indeed uniformly picked within
the allowed phase space, alternatively that any Jacobians are properly
taken into account. For instance, in the case above, one approach
would be to pick $x_1$, $x_2$ and $\hat{t}$ uniformly in the ranges
$0 < x_1 < 1$, $0 < x_2 < 1$, and $-s < \hat{t} < 0$, with full phase
space
volume $\int \d X = s$. The cross section would only be non-vanishing
inside the physical region given by $-s x_1 x_2 < \hat{t}$ (in the
massless case), i.e.\ Monte Carlo efficiency is likely to be low.
However, if one were to choose $\hat{t}$ values only in the range
$-\hat{s} < \hat{t} < 0$, small $\hat{s}$ values would be favoured,
since the density of selected $\hat{t}$ values would be larger there.
Without the use of a compensating Jacobian $\hat{s}/s$, an incorrect
answer would be obtained. Alternatively, one could start out with a
phase space like $\d X = \d x_1 \, \d x_2 \, \d (\cos\hat{\theta})$,
where the limits decouple. Of course, the $\cos\hat{\theta}$ variable
can be translated back into a $\hat{t}$, which will then always be in
the desired range $-\hat{s} < \hat{t} < 0$. The transformation itself
here gives the necessary Jacobian.
At times, it is convenient to split a process into a discrete set of
subprocesses for the parton-level generation, without retaining these
in the \ttt{IDPRUP} classification. For instance, the cross section above
contains a summation over incoming partons. An alternative would then
have been to let each subprocess correspond to one unique combination
of incoming flavours. When an event of process type $i$ is to be
generated, first a specific subprocess $ik$ is selected with probability
$f^{ik}$, where $\sum_k f^{ik} = 1$. For this subprocess an
\ttt{XWGTUP}$^k$ is generated as above, except that there is no longer
a summation over incoming flavours. Since only a fraction $f^{ik}$ of all
events now contain this part of the cross section, a compensating factor
$1/f^{ik}$ is introduced, i.e.\ \ttt{XWGTUP}$ = $\ttt{XWGTUP}$^k/f^{ik}$.
Further, one has to define
\ttt{XMAXUP(i)}$ = \mmax_k$ \ttt{XMAXUP}$^{ik}/f^{ik}$ and
\ttt{XSECUP(i)}$ = \sum_k$ \ttt{XSECUP}$^{ik}$.
The generation efficiency will be maximized for the $f^{ik}$ coefficients
selected proportional to \ttt{XMAXUP}$^{ik}$, but this is no requirement.
The standard allows external parton-level events to come with negative
weights, unlike the case for internal {\Py} processes. In order to avoid
indiscriminate use of this option, some words of caution are in place.
In next-to-leading-order calculations, events with negative weights
may occur as part of the virtual corrections. In any physical
observable quantity, the effect of such events should cancel against
the effect of real events with one more parton in the final state.
For instance, the next-to-leading order calculation of gluon scattering
contains the real process $\g\g \to \g\g\g$, with a positive divergence
in the soft and collinear regions, while the virtual corrections to
$\g\g \to \g\g$ are negatively divergent. Neglecting the problems of
outgoing gluons collinear with the beams, and those of soft gluons, two
nearby outgoing gluons in the $\g\g \to \g\g\g$ process can be combined
into one effective one, such that the divergences can be cancelled.
If rather widely separated gluons can be combined, the remaining
negative contributions are not particularly large. Different separation
criteria could be considered; one example would be
$\Delta R = \sqrt{(\Delta \eta)^2 + (\Delta \varphi)^2} \approx 1$.
The recombination of well separated partons is at the price of an
arbitrariness in the choice of clustering algorithm, when two gluons of
nonvanishing invariant mass are to be combined into one massless one,
as required to be able to compare with the kinematics of the massless
$\g\g \to \g\g$ process when performing the divergence cancellations.
Alternatively, if a smaller $\Delta R$ cut is used, where the combining
procedure is less critical, there will be more events with large positive
and negative weights that are to cancel.
Without proper care, this cancellation could easily be destroyed by
the subsequent showering description, as follows. The standard for
external processes does not provide any way to pass information on
the clustering algorithm used, so the showering routine will have to
make its own choice what region of phase space to populate with radiation.
One choice could be to allow a cone defined by the nearest
colour-connected parton (see section \ref{sss:showermatching} for a
discussion). There could then arise a significant mismatch in shower
description between events where two gluons are just below or just
above the $\Delta R$ cut for being recombined, equivalently between
$\g\g \to \g\g$ and $\g\g \to \g\g\g$ events. Most of the phase space
may be open for the former, while only the region below $\Delta R$ may
be it for the latter. Thus the average `two-parton' events may end up
containing significantly more jet activity than the corresponding
`three-parton' ones. The smaller the $\Delta R$ cut, the more severe
the mismatch, both on an event-by-event basis and in terms of the
event rates involved.
One solution would be to extend the standard also to specify which
clustering algorithm has been used in the matrix-element calculation,
and with what parameter values. Any shower emission that would give
rise to an extra jet, according to this algorithm, would be vetoed.
If a small $\Delta R$ cut is used, this is almost equivalent to allowing
no shower activity at all. (That would still leave us with potential
mismatch problems in the hadronization description. Fortunately the
string fragmentation approach is very powerful in this respect,
with a smooth transition between two almost parallel gluons and a
single one with the full energy \cite{Sjo84}.) But we know that the
unassisted matrix-element description cannot do a good job of the
internal structure of jets on an event-by-event basis, since
multiple-gluon emission is the norm in this region. Therefore a
$\Delta R \sim 1$ will be required, to let the matrix elements
describe the wide-angle emission and the showers the small-angle one.
This again suggests a picture with only a small contribution from
negative-weight events. In summary, the appearance of a large fraction
of negative-weight events should be a sure warning sign that physics
is likely to be incorrectly described.
The above example illustrates that it may, at times, be desirable to
sidestep the standard and provide further information directly in the
{\Py} common blocks. (Currently there is no exact match to the
clustering task mentioned above, although the already-described
\ttt{UPVETO} routine, section \ref{ss:PYwtevts}, could be constructed
to make use of such information. Here we concentrate on a few simpler
ways to intervene, possibly to be used in
conjunction with \ttt{UPVETO}.) Then it is useful to note that,
apart from the hard-process generation machinery itself, the external
processes are handled almost exactly as the internal ones. Thus
essentially all switches and parameter values related to showers,
underlying events and hadronization can be modified at will. This even
applies to alternative listing modes of events and history pointers,
as set by \ttt{MSTP(128)}. Also some of the information on the hard
scattering is available, such as \ttt{MSTI(3)},
\ttt{MSTI(21) - MSTI(26)}, and \ttt{PARI(33) - PARI(38)}. Before using
them, however, it is prudent to check that your variables of interest
do work as intended for the particular process you study. Several
differences do remain between internal and external processes, in
particular related to final-state showers and resonance decays. For
internal processes, the \ttt{PYRESD} routine will perform a shower
(if relevant) directly after each decay. A typical example would be
that a $\t \to \b\W$ decay is immediately followed by a shower,
which could change the momentum of the $\W$ before it decays in its
turn. For an external process, this decay chain would presumably
already have been carried out. When the equivalent shower to the
above is performed, it is therefore now necessary also to boost the
decay products of the $\W$. The special sequence of showers and boosts
for external processes is administrated by the \ttt{PYADSH} routine.
Should the decay chain not have been carried out, e.g if \ttt{HEPEUP}
event record contains an undecayed $\Z^0$, then \ttt{PYRESD} will
be called to let it decay. The decay products will be visible also
in the documentation section, as for internal processes.
You are free to make use of whatever tools you want in your
\ttt{UPINIT} and \ttt{UPEVNT} routines, and normally there would be
little or no contact with the rest of {\Py}, except as described above.
However, several {\Py} tools can be used, if you so wish. One
attractive possibility is to use \ttt{PYPDFU} for
parton-distribution-function evaluation. Other possible tools could
be \ttt{PYR} for random-number generation, \ttt{PYALPS} for $\alphas$
evaluation, \ttt{PYALEM} for evaluation of a running $\alphaem$, and
maybe a few more.
We end with a few comments on anomalous situations. As already
described, you may put \ttt{NUP = 0} inside \ttt{UPEVNT}, e.g.\
to signal the end of the file from which events are read.
If the program encounters this value at a return from \ttt{UPEVNT},
then it will also exit from \ttt{PYEVNT}, without incrementing the
counters for the number of events generated. It is then up to you to
have a check on this condition in your main event-generation loop.
This you do either by looking at \ttt{NUP} or at \ttt{MSTI(51)}; the
latter is set to 1 if no event was generated.
It may also happen that a parton-level configuration fails elsewhere
in the \ttt{PYEVNT} call. For instance, the beam-remnant treatment
occasionally encounters situations it cannot handle, wherefore the
parton-level event is rejected and a new one generated. This happens also
with ordinary (not user-defined) events, and usually comes about as a
consequence of the initial-state radiation description leaving too
little energy for the remnant. If the same hard scattering were to
be used as input for a new initial-state radiation and beam-remnant
attempt, it could then work fine. There is a possibility to give events
that chance, as follows. \ttt{MSTI(52)} counts the number of times a
hard-scattering configuration has failed to date. If you come in to
\ttt{UPEVNT} with \ttt{MSTI(52)} non-vanishing, this means that the
latest configuration failed. So long as the contents of the \ttt{HEPEUP}
common block are not changed, such an event may be given another try.
For instance, a line
\begin{verbatim}
IF(MSTI(52).GE.1.AND.MSTI(52).LE.4) RETURN
\end{verbatim}
at the beginning of \ttt{UPEVNT} will give each event up to five
tries; thereafter a new one would be generated as usual. Note that the
counter for the number of events is updated at each new try. The
fraction of failed configurations is given in the bottom line of
the \ttt{PYSTAT(1)} table.
The above comment only refers to very rare occurrences (less than
one in a hundred), which are not errors in a strict sense; for instance,
they do not produce any error messages on output. If you get
warnings and error messages that the program does not understand the
flavour codes or cannot reconstruct the colour flows, it is due to
faults of yours, and giving such events more tries is not going to
help.
\subsection{Interfaces to Other Generators}
\label{ss:PYinterfacegen}
In the previous section an approach to including external processes
in {\Py} was explained. While general enough, it may not always
be the optimal choice. In particular, for $\ee$ annihilation events
one may envisage some standard cases where simpler approaches could
be pursued. A few such standard interfaces are described in this
section.
In $\ee$ annihilation events, a convenient classification of electroweak
physics is by the number of fermions in the final state. Two fermions
from $\Z^0$ decay is LEP1 physics, four fermions can come e.g.\ from
$\W^+\W^-$ or $\Z^0\Z^0$ events at LEP2, and at higher energies six
fermions are produced by three-gauge-boson production or top-antitop.
Often interference terms are non-negligible, requiring much more complex
matrix-element expressions than are normally provided in {\Py}.
Dedicated electroweak generators often exist, however, and the task is
therefore to interface them to the generic parton showering and
hadronization machinery available in {\Py}. In the LEP2 workshop
\cite{Kno96} one possible strategy was outlined to allow reasonably
standardized interfaces between the electroweak and the QCD generators.
The \ttt{LU4FRM} routine was provided for the key four-fermion case. This
routine is now included here, in slightly modified form, together with
two siblings for two and six fermions. The former is trivial and
included mainly for completeness, while the latter is rather more delicate.
In final states with two or three quark--antiquark pairs, the colour
connection is not unique. For instance, a $\u\d\ubar\dbar$ final
state could either stem from a $\W^+\W^-$ or a $\Z^0\Z^0$ intermediate
state, or even from interference terms between the two. In order to
shower and fragment the system, it is then necessary to pick one of the
two alternatives, e.g.\ according to the relative matrix element weight
of each alternative, with the interference term dropped. Some different
such strategies are proposed as options below.
Note that here we discuss purely perturbative ambiguities. One can
imagine colour reconnection at later stages of the process, e.g.\ if
the intermediate state indeed is $\W^+\W^-$, a soft-gluon exchange
could still result in colour singlets $\u\ubar$ and $\d\dbar$. We are
then no longer speaking of ambiguities related to the hard process
itself but rather to the possibility of nonperturbative effects. This
is an interesting topic in itself, addressed in section
\ref{sss:reconnect} but not here.
The fermion-pair routines are not set up to handle QCD four-jet events,
i.e.\ events of the types $\q\qbar\g\g$ and $\q\qbar\q'\qbar'$ (with
$\q'\qbar'$ coming from a gluon branching). Such events are generated in
normal parton showers, but not necessarily at the right rate (a problem
that may be especially interesting for massive quarks like $\b$).
Therefore one would like to start a QCD final-state parton shower from
a given four-parton configuration. Already some time ago, a machinery
was developed to handle this kind of occurrences \cite{And98a}.
This approach has now been adapted to {\Py}, in a somewhat modified
form, see section \ref{sss:fourjetmatch}. The main change is that, in the
original work, the colour flow was
picked in a separate first step (not discussed in the publication, since
it is part of the standard 4-parton configuration machinery of \ttt{PYEEVT}),
which reduces the number of allowed $\q\qbar\g\g$ parton-shower histories.
In the current implementation, more geared towards completely external
generators, no colour flow assumptions are made, meaning a few more
possible shower histories to pick between. Another change is that mass
effects are better respected by the $z$ definition. The code contains one
new user routine, \ttt{PY4JET}, two new auxiliary ones, \ttt{PY4JTW} and
\ttt{PY4JTS}, and significant additions to the \ttt{PYSHOW} showering
routine.
\drawbox{CALL PY2FRM(IRAD,ITAU,ICOM)}\label{p:PY2FRM}
\begin{entry}
\itemc{Purpose:} to allow a parton shower to develop and partons to
hadronize from a two-fermion starting point. The initial list is
supposed to be ordered such that the fermion precedes the antifermion.
In addition, an arbitrary number of photons may be included, e.g.\ from
initial-state radiation; these will not be affected by the operation
and can be put anywhere. The scale for QCD (and QED) final-state radiation
is automatically set to be the mass of the fermion-antifermion pair.
(It is thus not suited for Bhabha scattering.)
\iteme{IRAD :} final-state QED radiation.
\begin{subentry}
\iteme{= 0 :} no final-state photon radiation, only QCD showers.
\iteme{= 1 :} photon radiation inside each final fermion pair, also leptons,
in addition to the QCD one for quarks.
\end{subentry}
\iteme{ITAU :} handling of $\tau$ lepton decay (where {\Py} does not
include spin effects, although some generators provide the helicity
information that would allow a more sophisticated modelling).
\begin{subentry}
\iteme{= 0 :} $\tau$'s are considered stable (and can therefore be decayed
afterwards).
\iteme{= 1 :} $\tau$'s are allowed to decay.
\end{subentry}
\iteme{ICOM :} place where information about the event (flavours,
momenta etc.) is stored at input and output.
\begin{subentry}
\iteme{= 0 :} in the \ttt{HEPEVT} common block (meaning that
information is automatically translated to \ttt{PYJETS} before
treatment and back afterwards).
\iteme{= 1 :} in the \ttt{PYJETS} common block. All fermions and photons
can be given with status code \ttt{K(I,1) = 1}, flavour code in
\ttt{K(I,2)} and five-momentum (momentum, energy, mass) in \ttt{P(I,J)}.
The \ttt{V} vector and remaining components in the \ttt{K} one are best
put to zero. Also remember to set the total number of entries \ttt{N}.
\end{subentry}
\end{entry}
\drawbox{CALL PY4FRM(ATOTSQ,A1SQ,A2SQ,ISTRAT,IRAD,ITAU,ICOM)}%
\label{p:PY4FRM}
\begin{entry}
\itemc{Purpose:} to allow a parton shower to develop and partons to
hadronize from a four-fermion starting point. The initial list of
fermions is supposed to be ordered in the sequence fermion (1) --
antifermion (2) -- fermion (3) -- antifermion (4). The flavour pairs
should be arranged so that, if possible, the first two could come
from a $\W^+$ and the second two from a $\W^-$; else each pair should
have flavours consistent with a $\Z^0$. In addition, an arbitrary number
of photons may be included, e.g.\ from initial-state radiation; these
will not be affected by the operation and can be put anywhere.
Since the colour flow need not be unique, three real and one
integer numbers are providing further input. Once the colour pairing
is determined, the scale for final-state QCD (and QED) radiation is
automatically set to be the mass of the respective fermion--antifermion
pair. (This is the relevant choice for normal fermion pair production
from resonance decay, but is not suited e.g.\ for $\gamma\gamma$ processes
dominated by small-$t$ propagators.) The pairing is also meaningful
for QED radiation, in the sense that a four-lepton final state is
subdivided into two radiating subsystems in the same way. Only if
the event consists of one lepton pair and one quark pair is the
information superfluous.
\iteme{ATOTSQ :} total squared amplitude for the event, irrespective of
colour flow.
\iteme{A1SQ :} squared amplitude for the configuration with fermions
$1 + 2$ and $3 + 4$ as the two colour singlets.
\iteme{A2SQ :} squared amplitude for the configuration with fermions
$1 + 4$ and $3 + 2$ as the two colour singlets.
\iteme{ISTRAT :} the choice of strategy to select either of the two
possible colour configurations. Here 0 is supposed to represent a
reasonable compromise, while 1 and 2 are selected so as to give the
largest reasonable spread one could imagine.
\begin{subentry}
\iteme{= 0 :} pick configurations according to relative probabilities
\ttt{A1SQ} : \ttt{A2SQ}.
\iteme{= 1 :} assign the interference contribution to maximize the
$1 + 2$ and $3 + 4$ pairing of fermions.
\iteme{= 2 :} assign the interference contribution to maximize the
$1 + 4$ and $3 + 2$ pairing of fermions.
\end{subentry}
\iteme{IRAD :} final-state QED radiation.
\begin{subentry}
\iteme{= 0 :} no final-state photon radiation, only QCD showers.
\iteme{= 1 :} photon radiation inside each final fermion pair, also leptons,
in addition to the QCD one for quarks.
\end{subentry}
\iteme{ITAU :} handling of $\tau$ lepton decay (where {\Py} does not
include spin effects, although some generators provide the helicity
information that would allow a more sophisticated modelling).
\begin{subentry}
\iteme{= 0 :} $\tau$'s are considered stable (and can therefore be decayed
afterwards).
\iteme{= 1 :} $\tau$'s are allowed to decay.
\end{subentry}
\iteme{ICOM :} place where information about the event (flavours,
momenta etc.) is stored at input and output.
\begin{subentry}
\iteme{= 0 :} in the \ttt{HEPEVT} common block (meaning that
information is automatically translated to \ttt{PYJETS} before
treatment and back afterwards).
\iteme{= 1 :} in the \ttt{PYJETS} common block. All fermions and photons
can be given with status code \ttt{K(I,1) = 1}, flavour code in
\ttt{K(I,2)} and five-momentum (momentum, energy, mass) in \ttt{P(I,J)}.
The \ttt{V} vector and remaining components in the \ttt{K} one are best
put to zero. Also remember to set the total number of entries \ttt{N}.
\end{subentry}
\itemc{Comment :} Also colour reconnection phenomena can be studied with the
\ttt{PY4FRM} routine. \ttt{MSTP(115)} can be used to switch between the
scenarios, with default being no reconnection. Other reconnection
parameters also work as normally, including that \ttt{MSTI(32)} can be
used to find out whether a reconnection occured or not. In order for the
reconnection machinery to work, the event record is automatically
complemented with information on the $\W^+ \W^-$ or $\Z^0 \Z^0$
pair that produced the four fermions, based on the rules described
above.\\
We remind that the four first parameters of the \ttt{PY4FRM} call are
supposed to parameterize an ambiguity on the perturbative level of the
process, which has to be resolved before parton showers are performed.
The colour reconnection discussed here is (in most scenarios) occuring
on the nonperturbative level, after the parton showers.
\end{entry}
\drawbox{CALL PY6FRM(P12,P13,P21,P23,P31,P32,PTOP,IRAD,ITAU,ICOM)}%
\label{p:PY6FRM}
\begin{entry}
\itemc{Purpose:} to allow a parton shower to develop and partons to hadronize
from a six-fermion starting point. The initial list of fermions is
supposed to be ordered in the sequence fermion (1) -- antifermion (2) --
fermion (3) -- antifermion (4) -- fermion (5) -- antifermion (6). The
flavour pairs should be arranged so that, if possible, the first two
could come from a $\Z^0$, the middle two from a $\W^+$ and the last two
from a $\W^-$; else each pair should have flavours consistent with a $\Z^0$.
Specifically, this means that in a $\t\tbar$ event, the $\t$ decay products
would be found in 1 ($\b$) and 3 and 4 (from the $\W^+$ decay) and the
$\tbar$ ones in 2 ($\bbar$) and 5 and 6 (from the $\W^-$ decay). In
addition, an
arbitrary number of photons may be included, e.g.\ from initial-state
radiation; these will not be affected by the operation and can be put
anywhere. Since the colour flow need not be unique, further input is
needed to specify this. The number of possible interference
contributions being much larger than for the four-fermion case, we
have not tried to implement different strategies. Instead six
probabilities may be input for the different pairings, that you
e.g.\ could pick as the six possible squared amplitudes, or according
to some more complicated scheme for how to handle the interference
terms. The treatment of final-state cascades must be quite different for
top events and the rest. For a normal three-boson event, each fermion
pair would form one radiating system, with scale set equal to the
fermion-antifermion invariant mass. (This is the relevant choice for
normal fermion pair production from resonance decay, but is not
suited e.g.\ for $\gamma\gamma$ processes dominated by small-$t$
propagators.)
In the top case, on the other hand, the $\b$ ($\bbar$) would be radiating
with a recoil taken by the $\W^+$ ($\W^-$) in such a way that the $\t$
($\tbar$) mass is preserved, while the $\W$ dipoles would radiate as normal.
Therefore you need also supply a probability for the event to
be a top one, again e.g.\ based on some squared amplitude.
\iteme{P12, P13, P21, P23, P31, P32 :} relative probabilities for the
six possible pairings of fermions with antifermions. The first (second)
digit tells which antifermion the first (second) fermion is paired with,
with the third pairing given by elimination. Thus e.g.\ \ttt{P23} means the
first fermion is paired with the second antifermion, the second fermion
with the third antifermion and the third fermion with the first
antifermion. Pairings are only possible between quarks and leptons
separately. The sum of probabilities for allowed pairings is
automatically normalized to unity.
\iteme{PTOP :} the probability that the configuration is a top one; a
number between 0 and 1. In this case, it is important that the order
described above is respected, with the $\b$ and $\bbar$ coming first.
No colour ambiguity exists if the top interpretation is selected,
so then the \ttt{P12 - P32} numbers are not used.
\iteme{IRAD :} final-state QED radiation.
\begin{subentry}
\iteme{= 0 :} no final-state photon radiation, only QCD showers.
\iteme{= 1 :} photon radiation inside each final fermion pair, also leptons,
in addition to the QCD one for quarks.
\end{subentry}
\iteme{ITAU :} handling of $\tau$ lepton decay (where {\Py} does not
include spin effects, although some generators provide the helicity
information that would allow a more sophisticated modelling).
\begin{subentry}
\iteme{= 0 :} $\tau$'s are considered stable (and can therefore be decayed
afterwards).
\iteme{= 1 :} $\tau$'s are allowed to decay.
\end{subentry}
\iteme{ICOM :} place where information about the event (flavours,
momenta etc.) is stored at input and output.
\begin{subentry}
\iteme{= 0 :} in the \ttt{HEPEVT} common block (meaning that
information is automatically translated to \ttt{PYJETS} before
treatment and back afterwards).
\iteme{= 1 :} in the \ttt{PYJETS} common block. All fermions and photons
can be given with status code \ttt{K(I,1) = 1}, flavour code in
\ttt{K(I,2)} and five-momentum (momentum, energy, mass) in \ttt{P(I,J)}.
The \ttt{V} vector and remaining components in the \ttt{K} one are best
put to zero. Also remember to set the total number of entries \ttt{N}.
\end{subentry}
\end{entry}
\drawbox{CALL PY4JET(PMAX,IRAD,ICOM)}\label{p:PY4JET}
\begin{entry}
\itemc{Purpose:} to allow a parton shower to develop and partons to
hadronize from a $\q\qbar\g\g$ or $\q\qbar\q'\qbar'$ original
configuration. The partons should be ordered exactly as indicated above,
with the primary $\q\qbar$ pair first and thereafter the two gluons or
the secondary $\q'\qbar'$ pair. (Strictly speaking, the definition of
primary and secondary fermion pair is ambiguous. In practice,
however, differences in topological variables like the pair mass
should make it feasible to have some sensible criterion on an
event-by-event basis.) Within each pair, fermion should precede
antifermion. In addition, an arbitrary number of photons may be included,
e.g.\ from initial-state radiation; these will not be affected by the
operation and can be put anywhere. The program will select a possible
parton-shower history from the given parton configuration, and then
continue the shower from there on. The history selected is displayed
in lines \ttt{NOLD + 1} to \ttt{NOLD + 6}, where \ttt{NOLD} is the \ttt{N}
value before the routine is called. Here the masses and energies of
intermediate partons are clearly displayed. The lines \ttt{NOLD + 7} and
\ttt{NOLD + 8} contain the equivalent on-mass-shell parton pair from which
the shower is started.
\iteme{PMAX :} the maximum mass scale (in GeV) from which the shower is
started in those branches that are not already fixed by the matrix-element
history. If \ttt{PMAX} is set zero (actually below \ttt{PARJ(82)}, the
shower cutoff scale), the shower starting scale is instead set to be equal
to the smallest mass of the virtual partons in the reconstructed
shower history. A fixed \ttt{PMAX} can thus be used to obtain a reasonably
exclusive set of four-jet events (to that \ttt{PMAX} scale), with little
five-jet contamination, while the \ttt{PMAX = 0} option gives a more
inclusive interpretation, with five- or more-jet events possible.
Note that the shower is based on evolution in mass, meaning the cut
is really one of mass, not of $\pT$, and that it may therefore be
advantageous to set up the matrix elements cuts accordingly if one
wishes to mix different event classes. This is not a requirement,
however.
\iteme{IRAD :} final-state QED radiation.
\begin{subentry}
\iteme{= 0 :} no final-state photon radiation, only QCD showers.
\iteme{= 1 :} photon radiation inside each final fermion pair, also leptons,
in addition to the QCD one for quarks.
\end{subentry}
\iteme{ICOM :} place where information about the event (flavours,
momenta etc.) is stored at input and output.
\begin{subentry}
\iteme{= 0 :} in the \ttt{HEPEVT} common block (meaning that
information is automatically translated to \ttt{PYJETS} before
treatment and back afterwards).
\iteme{= 1 :} in the \ttt{PYJETS} common block. All fermions and photons
can be given with status code \ttt{K(I,1) = 1}, flavour code in
\ttt{K(I,2)} and five-momentum (momentum, energy, mass) in \ttt{P(I,J)}.
The \ttt{V} vector and remaining components in the \ttt{K} one are best
put to zero. Also remember to set the total number of entries \ttt{N}.
\end{subentry}
\end{entry}
\subsection{Other Routines and Common Blocks}
The subroutines and common blocks that you will come in direct
contact with have already been described. A number of other routines
and common blocks exist, and those not described elsewhere are here
briefly listed for the sake of co