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\begin{document}
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%\preprint{TTP15-009}
\title{
\vskip-1cm{\baselineskip14pt
\begin{flushright}
\normalsize TTP15-009
\end{flushright}}
% \vskip1.5cm
Production of two Z-bosons in gluon fusion
in the heavy top quark approximation
}
\author{Kirill Melnikov}
\email{melnikov@pha.jhu.edu}
\affiliation{ Institute for Theoretical Particle Physics, Karlsruhe Institute of Technology,
Karlsruhe, Germany}
%Department of Physics and Astronomy, Johns Hopkins University, Baltimore, USA
%}
\author{Matthew Dowling }
\email{matthew.dowling@kit.edu}
\affiliation{ Institute for Theoretical Particle Physics, Karlsruhe Institute of Technology,
Karlsruhe, Germany}
%Department of Physics and Astronomy, Johns Hopkins University, Baltimore, USA
%}
\begin{abstract}
We compute QCD radiative corrections to the continuum production of
a pair of $Z$-bosons in the annihilation of two gluons. We only consider
the contribution of the top quark loops and we treat them assuming that
$m_t$ is much larger than any other kinematic invariant in the problem.
We estimate the QCD corrections to $pp \to ZZ$ using the first non-trivial term in the expansion
in the inverse top quark mass and we compare them to QCD corrections of the
signal process, $pp \to H \to ZZ$.
\end{abstract}
\maketitle
\section{Introduction}
Production of pairs of vector bosons in proton collisions is one of the most interesting processes
studied at the LHC at the end of the Run I \cite{atlas7,cms7,cms8}.
Indeed, $pp \to ZZ$, $pp \to W^+W^-$,
and $pp \to \gamma \gamma$, play an important role in Higgs boson physics, provide stringent tests
of the Standard Model and give constraints on anomalous electroweak triple gauge boson couplings.
In the case of Higgs physics, such processes are essential for understanding backgrounds to Higgs
boson signals, for constraining anomalous Higgs boson couplings, for measuring the
quantum numbers of the Higgs boson and for studying the Higgs boson width, see e.g.
Refs.\cite{Khachatryan:2014jba,Khachatryan:2014kca,Khachatryan:2014iha,atlaswidth}.
Production of electroweak gauge boson pairs occurs mainly due to quark-antiquark annihilation.
This contribution is known through next-to-next-to-leading order (NNLO) in perturbative QCD
\cite{Cascioli:2014yka,tg1}. However, as was pointed out in
Refs.\cite{Glover:1988rg,Glover:1988fe,Dicus:1987dj},
there is a sizeable contribution of the gluon annihilation channel
$gg \to V_1 V_2$ whose significance depends on the selection cuts. For example ~\cite{Binoth:2006mf},
agressive cuts applied to $pp \to W^+W^-$ to separate the Higgs boson signal from the continuum
background can increase
the fraction of gluon fusion events in the background sample to ${\cal O}(30)$ percent.
Since $gg \to V_1 V_2$ is the one-loop process and since production of
electroweak boson pairs at leading order occurs only in the $q \bar q$ channel, the gluon fusion contribution
to $pp \to V_1 V_2$ through NNLO only needs to be known at the leading, one-loop, approximation.
Thus, all existing numerical estimates of the significance of the
gluon fusion mechanism in weak boson pair production ignore radiative corrections
to $gg \to ZZ$ that are, potentially, quite large~\cite{Bonvini:2013jha}. The need
to have a more accurate estimate of QCD corrections to gluon fusion processes was strongly
emphasized in Ref.~\cite{atlaswidth}, in the context of the Higgs width and generic off-shell
measurements \cite{Caola:2013yja,
Campbell:2013una, ellis,Azatov:2014jga}.
The largest contribution to $gg \to V_1 V_2$ comes from quarks of the first two generations
that can be taken to be massless (for a recent discussion, see Ref.~\cite{ellis}).
The contribution of the third generation is, in general
smaller. For example, in the case of $W^+W^-$ production it is known that the third generation
changes the $gg \to V_1 V_2$ production cross-section by ${\cal O}(10)$ percent at the
$13~{\rm TeV}$ LHC \cite{ellis}. Since gluon fusion contributes ${\cal O}(5)$ percent
to the $pp \to W^+W^-$ cross-section, the impact of top quark loops on the cross-section is marginal.
On the other hand, studies of {\it off-shell}
Higgs boson production may be senstitive to the third
generation of quarks and, especially, to {\it massive top quark loops}. Of particular concern in this context
is the interference of $gg \to V_1 V_2$ and $gg \to H^* \to V_1 V_2$ amplitudes,
as discussed recently in Refs.~\cite{ellis,Azatov:2014jga}.
The recent progress in calculating two-loop integrals with two massless and two massive external lines
\cite{Gehrmann:2014bfa,nonplanar,planar,Papadopoulos:2014hla}
enables computation of scattering amplitudes and, eventually, QCD corrections
to the production of pairs of vector bosons in gluon fusion through loops of {\it massless} quarks.
A similar progress towards computing $gg \to V_1 V_2$ contributions mediated by loops of massive
quarks is very desirable but, probably, it will not be immediate
since two-loop computations of four-point functions with internal massive lines are beyond
existing technical capabilities.
In this situation, it is useful to think about alternative approaches that will allow
an estimate of QCD radiative corrections to gluon fusion processes
mediated by heavy quark loops.
A practical opportunity is provided by the expansion of amplitudes in the inverse quark mass.
Indeed, this approach reduces the calculation of the one-loop $gg \to ZZ$ amplitude
with massive internal particles to the calculation of tadpole diagrams which makes generalization
to higher-order corrections relatively straightforward.\footnote{A similar approach is more problematic
in the case of $gg \to W^+W^-$ where the third generation loops contain both massive (top) and massless (bottom) quarks.}
While the expansion of cross-sections in $1/m_t$
cannot be fully justified, particularly for large invariant masses of $Z$-pairs, we have significant
evidence that such computations do provide a reasonable estimate of the size of QCD corrections.
Indeed, this is an approach that is taken in calculations of single- \cite{shiggs} and double-Higgs
\cite{dhiggs} production
at the LHC where {\it exact} one-loop computations supplemented with QCD corrections
calculated in the $m_t \to \infty$ approximation are believed to provide reasonably accurate descriptions
of these processes for realistic values
of top quark and Higgs boson masses.
It is clear that a similar approach should be applicable to the production of pairs of
$Z$- bosons in gluon fusion through the top quark loop. In fact, the $m_t \to \infty$
approximation for $gg \to ZZ$ should work better than for the case of Higgs pair production
since $2m_Z$ is smaller than $2 m_H$.
The goal of this paper is to make the first step
towards estimating the NLO QCD correction to the production of $Z$-boson pairs in gluon
fusion. To this end, we take {\it continuum} production of two on-shell $Z$-bosons in gluon fusion
through the top quark loop and compute the NLO QCD corrections to it in the heavy top approximation.
This allows us to compare, for the first time, the QCD corrections to the ``background'' $gg \to ZZ$ and
the signal $gg \to H \to ZZ$ processes. We find that the corrections to the two processes are
indeed similar, in accord with the arguments in Ref.~\cite{Bonvini:2013jha}.
It should be clear from the previous discussion that computation of QCD corrections to
the total cross-section is just one of many
interesting physics questions,
including
interference with the Higgs signal on and off the mass shell,
combination of light and heavy quark contirbutions to the
$gg \to ZZ$ amplitude, estimates of $1/m_t$ corrections to
cross-sections etc., that can be discussed in the context of vector
boson pair production in gluon fusion,
once the two-loop amplitude for $gg \to ZZ$ becomes available.
We plan to address these questions in the near future.
The paper is organized as follows. In Section~\ref{section1}, we describe the general
set up of the computation and present the analytic result for the two-loop amplitude $gg \to ZZ$
in the large-$m_t$ approximation.
In Section~\ref{section2}, we derive the analytic formulas for
$gg \to ZZ$ partonic cross-sections. In Section~\ref{section3} we discuss numerical results.
We present our conclusions in Section~\ref{section4}.
\section{The set up of the computation}
\label{section1}
We consider the process $g(p_1) + g(p_2) \to Z(p_3) + Z(p_4)$ in a theory where $Z$-bosons
only couple to top quarks.\footnote{Such a theory is anomalous and, in principle, one should carefully
consider diagrams where a $Z$-boson couples to {\it two} gluons. Given the order in the $1/m_t$
expansion that we work to in this paper, Feynman diagrams where each $Z$ independently couples to
gluon pairs do not contribute.}
Contributions of massless quarks are not included in the computation
except in the running of the coupling constant where the complete $\beta$-function is employed.
The coupling of top quarks to $Z$-bosons is given by a linear combination
of vector and axial couplings
\be
Z \bar t t \in -i \gamma^\mu ( g_V + g_A \gamma_5),
\ee
where $g_V = e/(2 \sin 2 \theta_W) ( 1- 8/3 \sin^2 \theta_W) $ and
$g_A = e/(2 \sin 2 \theta_W)$.
\begin{figure}[t]
\centering
\includegraphics[angle=0,width=0.4\textwidth]{ggVVt.eps}\\
\caption{Representative two-loop diagrams that describe production of Z-boson pairs
in gluon fusion. }
\label{fig0}
\end{figure}
The scattering amplitude for $gg \to ZZ$ can be written as a sum of axial, vector and
mixed terms
\be
{\cal A}_{gg \to ZZ} = i a_s \delta^{a_1 a_2}
\left ( g_A^2 {\cal A}^{aa} + g_V^2 {\cal A}^{vv} + g_A g_V {\cal A}^{av} \right ),
\label{eq1}
\ee
where $a_{1,2}$ are the color indices of the colliding gluons,
$$
a_s = \frac{\Gamma(1+\ep)}{(4\pi)^{-\ep}} \; \frac{ \alpha_s(\mu)}{\pi },
$$
and $\alpha_s(\mu)$ is the $\overline {\rm MS}$ QCD coupling constant in the theory with five active
flavors.\footnote{ The contributions to the running of the coupling constant due to top quarks
are subtracted at zero external momentum, i.e. on the mass-shell of an external gluon.}
We note that thanks to charge parity conservation, the axial-vector term vanishes, i.e. $A^{av} = 0$.
The remaining two terms -- axial-axial ${\cal A}^{aa} $
and vector-vector ${\cal A}^{vv}$ -- do not vanish but they
behave differently under the $1/m_t$ expansion.
Indeed, consider a vector-current interaction of $Z$-bosons with top quarks.
The $gg \to ZZ$ amplitude behaves as ${\cal A}^{vv} \sim s^2/m_t^4$. This is a direct consequence
of the vector current conservation which requires that, in the
expression for the amplitude, each polarization vector for
either a gluon or an electroweak gauge boson is accompanied by its momentum. A similar suppression
in the QED case is familiar in the context of Euler-Heisenberg Lagrangian.
However, if the interaction of $Z$-bosons with top quarks is mediated
by the axial current, the situation is different since the axial current is not conserved.
As a consequence, the scattering amplitude can involve two polarization vectors of the
$Z$-bosons {\it without} corresponding momenta while the gluon polarization vectors should still
be accompanied by their momenta to satisfy the vector current conservation constraint. Hence,
we expect that the axial amplitude ${\cal A}^{aa}$ behaves as ${\cal A}^{aa} \sim s/m_t^2$ and, therefore,
exhibits weaker suppression in the $m_t \to \infty$
limit compared to ${\cal A}^{vv}$.
Since our goal in this paper is to study the leading term of the ${\cal A}_{gg \to ZZ}$ amplitude
in the $m_t \to \infty$ expansion, we conclude that we only need to study terms induced by the axial
coupling of the $Z$-bosons to top quarks.
The production of $Z$-boson pairs in gluon fusion is a loop-induced process.
There are eight one-loop and ninty-three two-loop diagrams that contribute
to $gg \to ZZ$. Some examples are shown in Fig.~\ref{fig0}. We compute these diagrams
using asymptotic expansions in the inverse top quark mass \cite{Smirnov:2013}. The essense
of this procedure is that the loop momenta in each of the Feynman diagrams
are separated into soft $l \sim p_{1,..4}$ and hard $l \sim m_t$. All possible
assignments must be considered. The integrand of a Feynman diagram is then Taylor
expanded in all quantities that are considered small. Upon such an expansion, computation of Feynman
diagrams significantly simplifies. Consider one-loop diagrams as an example.
In this case the momentum can only be hard,
$l \sim m_t$,
and so integrands for
all diagrams are expanded in Taylor series in their external momenta.\footnote{If the loop momentum
is assumed to be soft, each propagator is expanded in $l/m_t$ generating scaleless integrals.}
All one-loop integrals then become vacuum tadpole integrals and it is straightforward to
evaluate them.
The situation with two-loop integrals is similar although somewhat more involved.
Indeed, in this case two momentum configurations
are possible: either both loop momenta are
hard or one of the loop momenta is hard and the other one is soft. If both loop momenta
are hard, the calculation is reduced to the calculation of two-loop vacuum tadpole
diagrams. If one of the loop momenta is soft and the other one is hard, the diagram
factorizes into a product of one-loop integrals, the most complicated
of which is a three-point function with all internal and two external lines massless.
We will now present our results for the $gg \to ZZ$ amplitude.
We write it
as an expansion in the strong coupling constant
\be
{\cal A}^{aa} = \frac{1}{3 m_t^2} \left(\frac{\mu}{m_t} \right )^{2\ep}
\left \{ {\cal A}^{aa}_{1} + a_s \; \left(\frac{\mu}{m_t} \right )^{2\ep} {\cal A}^{aa}_{2} \right \}.
\ee
To emphasize constraints on the amplitude that follow from
gauge invariance, we introduce the Fourier transform of the
field-strength tensor for each of the gluons
\be
f^{i,\mu \nu} = p_i^{\mu} \epsilon_i^{\nu} - p_i^{\nu} \epsilon_i^{\mu},\;\;\;\;i =1,2.
\ee
The one-loop amplitude reads
\be
\begin{split}
{\cal A}^{aa}_{1} = (1+\ep) \left ( f_{\mu \rho}^{1} f^{2,\mu}_{\beta}
- \frac{g_{\rho \beta}}{2} f_{\mu \rho}^{1} f^{2,\mu \rho} \right ) t_{34}^{\rho \beta},
\label{eq7}
\end{split}
\ee
where
\be
t_{34}^{\rho \beta} = \epsilon_3^{\rho} \epsilon_{4}^{\beta}
+ \epsilon_{4}^{ \rho} \epsilon_{3}^{\beta},
\ee
and $\epsilon_{3,4}$ are the polarization vectors of the two $Z$-bosons.
We emphasize that the dependence of the amplitude on the dimensional regularization
parameter $\ep$ in Eq.(\ref{eq7}) is exact.
The two-loop amplitude reads
\be
\begin{split}
{\cal A}^{aa,2} & = \left (
- \left ( \frac{3}{2\ep^2} + \frac{\beta_0}{2 \ep} \right ) \left ( \frac{-s-i0}{m_t^2} \right )^{-\ep}
\right.
\\
& \left. -\frac{\beta_0}{2} L_{s\mu} + \frac{11}{4} L_{sm} + \frac{\pi^2}{4} - \frac{175}{36}
\right ) {\cal A}^{aa,1}
\\
&
+\frac{1}{2} f_{\mu \rho}^{1} f^{2,\mu \rho} t_{34 \beta}^{\beta} \left (-\frac{385}{72}
+ \frac{11}{8} L_{sm} \right )
\\
& - \frac{1}{2s} f_{\mu \nu}^{1} f^{2,\mu \nu} t_{34}^{\rho, \beta} (p_{1,\rho}p_{1,\beta} + p_{2,\rho} p_{2, \beta} )
\\
&
+ \frac{3}{2s} f_{\mu \rho}^{1} p_1^\mu f^{2}_{\nu \beta} p_2^{\nu} t_{34}^{\rho, \beta} + \mathcal{O}(\epsilon),
\end{split}
\ee
where $L_{s\mu} = \log((-s-i0)/\mu^2)$ and $L_{sm} = \log((-s-i0)/m_t^2)$ and $\beta_0 = 11/2 - N_f/3 $ with $N_f=5$
being the number of {\it massless} fermions.
In addition to virtual corrections, we require an amplitude for the real emission process,
$g(p_1)g(p_2) \to Z(p_3) + Z(p_4) + g(p_5) $.
To order $1/m_t^2$ the corresponding amplitude can, in principle, be obtained from the amplitude of the one-loop scattering
process in Eq.(\ref{eq7}), if the latter is written as a term in an effective Lagrangian, and then used to generate
amplitudes with additional gluons in the final state.
However,
it is also convenient to apply the asymptotic expansion procedure to the computation of the relevant
diagrams since this approach can be used to obtain the amplitude for $gg \to ZZ+g$ beyond the leading
order in $1/m_t$.
We use the second approach to compute the $gg \to ZZ+g$ amplitude. There are fifty diagrams that
contribute to this process and we compute the relevant diagrams using the $1/m_t$ expansion. The technical
details of the calculation are identical to the calculation of the scattering amplitude for the
$gg \to ZZ$ process and we do not repeat it here. Unfortunately, the expresssion for the amplitude
appears to be too complex to be presented here.
To calculate the production cross-section, we square the elastic and the inelastic scattering
amplitudes and integrate them over the corresponding phase-spaces. In order to make the cross-section
finite, we need to remove collinear singularities by performing renormalization of parton distribution
functions. All of these steps are relatively standard and well-known; for this reason we refrain from
describing them in detail.
\section{Production cross-section}
\label{section2}
We are now in position to present results for the gluon fusion contribution
to the production cross-section $pp \to ZZ$. As explained previously, we only
consider loops of top quarks and we work to leading order in the $1/m_t$ expansion.
We take the invaraint mass of the $Z$-boson pair to be $q^2$ and write the
differential cross-section as
a convolution of the partonic production
cross-section and the parton distribution functions
\be
\begin{split}
\frac{{\rm d} \sigma_{pp \to ZZ} }{{\rm d} q^2} & =
\int \limits_{0}^{1} {\rm d} x_1 {\rm d} x_2 {\rm d} z \; f_g(x_1) f_g(x_2)
\\
&
\times \delta\left ( z - \frac{\tau}{x_1 x_2} \right )
\frac{{\rm d} \sigma_{gg \to ZZ}}{{\rm d } q^2}(s,q^2)|_{s = q^2/z}.
\end{split}
\label{eq8}
\ee
In Eq.(\ref{eq8}), we used the following notation: $f_{g}(x_{1,2})$ are the
gluon parton distribution functions, $\tau = q^2/S_{\rm hadr}$ and
$S_{\rm hadr}$ is the hadronic center-of-mass energy squared.
We note that dependencies on the renormalization
and factorization scales in Eq.(\ref{eq8}) are suppressed.
In what follows, we take the factorization and the renormalization scales
to be equal.
It is conventional to parametrize
the partonic cross-section as
\be
q^2\frac{{\rm d} \sigma_{gg \to ZZ}}{{\rm d } q^2}(s,q^2)|_{s = q^2/z} = \sigma_0 z G(z,q^2),
\ee
where
\be
\sigma_0 = \frac{g_A^4 q^2 }{2^{10} \pi m_t^4}
\left ( \frac{\alpha_s(\mu)}{\pi} \right )^2 \sqrt{1-\frac{4m_Z^2}{q^2}},
\ee
and $G(z,q^2)$ can be written as series in the strong coupling constant. To present it, we
introduce a parameter $r$ defined as $r = q^2/(4m_Z^2)$. We find
\be
\begin{split}
& G(z,q^2) =
\Bigg [ \Delta_0 \delta(1-z)
+ a_s \Big (
\Delta_V \delta(1-z) +
\\
&
6\Delta_0 \left ( 2D_1(z) +\ln \frac{q^2}{\mu^2} D_0(z) \right )
+ \Delta_H
\Big ) \Bigg ],
\end{split}
\ee
where $D_i(z) = \left [ \ln(1-z)^i/(1-z) \right ]_+$ are the different plus-distribution functions and
\be
\label{eqdelta0}
\begin{split}
& \Delta_0 = \frac{73}{270} - \frac{2r}{15} + \frac{34 r^2}{135}.
\end{split}
\ee
We note that $\Delta_0$ has a strong dependence on $q^2$. The leading growth caused by the ${\cal O}(r^2) \sim q^4/m_Z^4$
term in Eq.(\ref{eqdelta0})
is the consequence of the fact that pairs of longitudinal bosons can be produced. It is this growth
that should, eventually, get tamed by the destructive interference of $gg \to ZZ$ and $gg \to H^* \to ZZ$
amplitudes.
The virtual corrections combined with finite parts of soft emissions read
\be
\begin{split}
& \Delta_V = \frac{2473 - 8661 r + 5798 r^2}{2430}
\\
& + \frac{ ( 73 - 36 r + 68 r^2 ) \pi^2}{270}
+ \frac{11(7+6r + 2r^2)}{135} \ln \frac{q^2}{m_t^2}.
\end{split}
\ee
The contributions of hard emissions, not proportional to the leading order cross-section read
\be
\begin{split}
\Delta_H & =
\frac{6\Delta_0}{z}
\left ( (\omega(z) - z\kappa(z))
\ln
\left ( \frac{q^2 (1-z)^2 }{\mu^2 } \right )
\right.
\\
& \left. - \omega(z)^2 \frac{\ln(z)}{(1-z)} \right )
+ (1-z) \Bigg [ \frac{r (11 \kappa(z) -46 z)}{15 z}
\\
&
- \frac{r^2 (187 \kappa(z) -302z)}{135 z}- \frac{(803 \kappa(z)-598z)}{540 z}
\Bigg ],
\end{split}
\ee
where $\omega(z) = 1-z+z^2$ and $\kappa(z) = 1+z^2$.
\begin{figure}[t]
\centering
\includegraphics[angle=0,width=0.45\textwidth]{K-factor.eps}\;\;
% \includegraphics[angle=0,width=0.22\textwidth]{HK-factor.pdf}\\
\caption{Main plot: NLO $K$-factor for $gg \to ZZ$ production through the top quark loop
as a function of the invariant mass of the $Z$-boson pair $q$, in GeV. Inset:
NLO $K$-factor for $gg \to H$ as a function of the Higgs boson mass $q$, in GeV.
Bands correspond to variations of the renormalization and factorization scales in the interval
$ q/4 \le \mu \le q$. The dashed line shows the $K$-factors computed for the renormalization
and factorization scales set to $\mu = q/2$.
We used the program MCFM \cite{mcfm} to compute the $K$-factor for the Higgs boson
production.
}
\label{fig1}
\end{figure}
\section{Numerical results}
\label{section3}
We have implemented the above formulas in a numerical {\sf Fortran} program that allows us to compute QCD corrections
to the top quark loop contribution to the gluon fusion process
$pp \to ZZ$ as a function of the invariant mass of the $Z$-bosons, $q^2$. We employ NNPDF3.0 parton
distribution functions \cite{Ball:2014uwa}
and use leading order parton distributions to compute the production cross-section at
leading (one-loop) approximation and next-to-leading order parton distirbutions to calculate it in the two-loop approximation.
We set the renormalization and factorization scales equal to each other.
To assess the magnitude of QCD corrections, in the main plot of Fig.~\ref{fig1} we show the $K$-factor defined as the
ratio of NLO and LO
cross-sections, depending on the invariant mass of the $Z$-boson pair. We find that the $K$-factor
is a slowly rising function of $q^2$ and that $K \sim 1.5-1.8$ for $\mu = q/2$ for the invariant masses
considered. The NLO QCD corrections to the $gg \to ZZ$ process are therefore similar to what has been
observed for other processes where gluons annihilate into colorless final states. As an illustration,
we compare the above results with the NLO $K$-factors for Higgs boson production $pp \to H$,
shown in the inset of Fig.~\ref{fig1}. We take the Higgs bosons
mass to be equal to the invariant mass of the $Z$-boson pair. The $K$-factors disagree by about
$10-15$ percent at low values of $q^2$ and agree almost perfectly at high(er) values of $q^2$.
This is in accord with the suggestion of Ref.~\cite{Bonvini:2013jha} where it was proposed
to employ the signal $K$-factor for the description of the complete process $gg \to ZZ$ including the continuum
contribution.
\begin{figure}[t]
\centering
\includegraphics[angle=0,width=0.4\textwidth]{MCFM173.eps}\\
\caption{LO $pp \to ZZ$ production cross-section ( gluon fusion through top loop only)
${\rm d}\sigma/{\rm d} q^2$ in ${\rm fb/GeV}^2$,
as a function of the invariant mass squared of the $Z$-boson pair, $q^2$,
in GeV${}^2$ with a top mass of 173 GeV.
We compare our cross-section, which is valid in the $m_t \to \infty$ limit, with the one implemented in MCFM, which has exact $m_t$ dependence.
The dots and squares correspond to the results from MCFM and this paper respectively.
We set the renormalization scale and the factorization scale to $200~\mathrm{GeV}$.
The difference between the values ranges from $\sim 20\%$ to $\sim 220\%$ for the values of $q^2$ considered.
The inset shows the MCFM cross-section computed for a larger range of the invariant masses.
}
\label{fig3}
\end{figure}
Finally, it is interesting to assess how the cross-section expanded in
$1/m_t$ compares with
the exact result. To this end, we show
in Fig.~\ref{fig3} the leading order contribution to
$gg \to ZZ$ with exact dependence on $m_t$ and the $m_t \to \infty$ limit. The exact
result is obtained using the program MCFM~\cite{mcfm}.
We see that near threshold values of $q^2$, the two cross-sections are similar, within $\sim 20\%$.
At larger values of $q^2$ the predictions start to diverge as $1/m_t$ supressed terms become more important.
As a check of our LO cross-section we performed a similar comparison with MCFM using a top mass
of $m_t = 400~\mathrm{GeV}$ to simulate the $m_t \to \infty$ limit.
In this case, our calculation is within $\sim 5\%$ of the MCFM predictions for values
of $q^2$ between $(200~\mathrm{GeV})^2$ and $(400~\mathrm{GeV})^2$. The inset in
Fig.~\ref{fig3} shows the top quark loop contribution to the
$pp \to ZZ$ cross-section with full mass dependence, as obtained with MCFM.
The cross-section peaks slightly above $400~{\rm GeV}$ and then starts to decrease. Our $K$-factor
calculation is valid to the left of the peak, where the cross-section exhibits rapid growth.
However, it can be extended beyond that by reweighting the {\it exact} $gg \to ZZ$ leading order
partonic cross-section with $K$-factors computed in $m_t \to \infty$ limit.
\section{Conclusions}
\label{section4}
In this paper we studied QCD corrections to
the production of a pair of $Z$-bosons in gluon fusion through loops of massive
top quarks. This process occurs at one-loop and belongs to the class of processes
where two gluons annihilate into a colorless final state. Similar to other processes
of this type, such as $gg \to H$ and $gg \to HH$, we find large, ${\cal O}(50-100)$ percent
radiative corrections. Radiative corrections of this magnitude suggest
that the significance of gluon fusion is, perhaps, underestimated by existing NNLO QCD computations of vector
boson pair production in proton collisions.
There are several avenues that are interesting to explore as the direct continuation of this work.
First, a straightforward extension of this calculation should allow us to compute QCD radiative
corrections to the interference
of $gg \to ZZ$ and $gg \to H \to ZZ$ amplitudes, both on and off the mass shell of the Higgs boson.
Although such a computation will, at the moment, be restricted to top quark contributions to
$gg \to ZZ$, it will already give us important information on whether or not
the radiative corrections to $gg \to H \to ZZ$, $gg \to ZZ$ and the interference are related.
Second, it will be interesting to extend our calculation to include higher powers in the
expansion of $s/m_t^2$, to estimate the impact of mass suppressed effects on QCD radiative
corrections. In addition, as we explained at the beginning of the paper, the effects of the
vector coupling of $Z$-bosons to top quarks do not appear at leading order in the $1/m_t$ expansion which
means that it is important to go one order higher in $1/m_t$ to fully incorporate physics of $Zt\bar t$
interactions into the description of the process. Of course, it is to be expected that
since the vector coupling of $Z$-bosons to
top quarks is almost three times smaller than the axial coupling, the inclusion of the
vector coupling should only lead to small changes in the
cross-section.
Third, it is interesting to incorporate decays of $Z$-bosons, off-shell effects
and realistic selection criteria into our calculation. This should, in principle, be straightforward
since the primary objects that we compute are the scattering amplitudes for both $gg \to ZZ^*$ and
$gg \to ZZ^*+g$ processes.
Finally, it is important to combine contributions of massless quarks and
the top quark to $gg \to ZZ$ amplitudes in higher orders of QCD. Since the two-loop amplitudes
for $gg \to ZZ$ with massless intermediate quarks are within reach
\cite{Caola:2014iua,Cascioli:2014yka}, it should
be relatively straightforward to incorporate both massless and massive quark
loops into the description of gluon fusion contribuitions to $Z$-boson pair production.
To conclude, we described the calculation of NLO QCD corrections to continuum production
of $Z$-boson pairs
in gluon fusion. The results of this computation present first {\it direct} evidenece that the
gluon fusion production cross-section of $Z$-bosons receives large ${\cal O}(100\%)$ QCD radiative corrections.
Our result should encourage further studies of QCD radiative corrections to weak boson pair production
in gluon fusion processes, mediated by massless and massive quark loops.
{\bf Acknowledgements}
K.M. would like to thank K.~Chetyrkin for useful conversations.
This research is partially supported by Karlsruhe Institute of Technology through its startup grant.
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Since it is most expensive to compute the form-factors, it is advantegious to use same form-factors that are computed
for the $ u \bar u$ amplitude. This requires that we change the parametrization of momenta and
consider the amplitude for the following process $d (p_2) \bar d(p_1) \to W^+(p_3) W^-(p_4)$.
The amplitude is obtained from the previous one by making the substitution $1 \leftrightarrow 2$,
$ 5 \leftrightarrow 7$ and $6 \leftrightarrow 8$. Note that these changes imply $3 \leftrightarrow 4$.
We obtain
\be
\begin{split}
A^{(d)}_{-} = &
- F_1 \langle 7 5\rangle [ 6 8] \langle 1 \hat 4 2]
+ F_2 \langle 2 7\rangle \langle 2 5\rangle [ 2 8 ][ 2 6]
\langle 1 \hat 4 2]
+ F_3 \langle 2 7\rangle \langle 1 5\rangle [ 28][ 1 6]
\langle 1 \hat 4 2]
\\
& + F_5 \langle 2 5\rangle \langle 1 7\rangle [ 2 6] [ 1 8]
\langle 1 \hat 4 2]
+ F_6 \langle 1 7\rangle \langle 1 5\rangle [ 1 8] [ 1 6]
\langle 1 \hat 4 2]
+ F_{14} \langle 2 7\rangle \langle 1 5\rangle [ 2 8] [ 2 6]
\\
&
+ F_{11} \langle 1 7 \langle 2 5\rangle \rangle [ 2 8][ 2 6]
+ F_{12} \langle 1 7\rangle \langle 1 5\rangle [ 2 8][ 1 6]
+ F_{15} \langle 1 7\rangle \langle 1 5\rangle [ 1 8] [ 2 6 ],
\\
A^{(\bar d )}_{+} = &
-F_1 \langle 75 \rangle [68] [1 \hat 4 2 \rangle
+ F_2 \langle 2 7 \rangle \langle 2 5 \rangle [2 8] [2 6] [1 \hat 4 2 \rangle
+ F_3 \langle 2 7\rangle \langle 1 5\rangle [ 28][ 1 6]
\langle 1 \hat 4 2]
\\
& + F_5 \langle 2 5\rangle \langle 1 7\rangle [ 2 6] [ 1 8]
\langle 1 \hat 4 2]
+ F_6 \langle 1 7\rangle \langle 1 5\rangle [ 1 8] [ 1 6]
\langle 1 \hat 4 2]
+ F_{14} [16] \langle 27 \rangle [28] \langle 25 \rangle
\\
& +F_{11} [18] \langle 27 \rangle [25] [26]
+ F_{12} [16] \langle 27 \rangle [18] \langle 15 \rangle
+F_{15} [16] \langle 17 \rangle [18] [25],
\end{split}
\ee
Finally, we give the expressions for the part of the amplitude moderated by photon and $Z$-boson exchanges.
It reads
\be
\begin{split}
& A_{-}^{Z/\gamma} = 4 F_V(s) \left ( 2 \sin^2 \theta_W \; Q_{u} D_{\gamma}(s) + (V_{u} + A_{u}) D_Z(s) \right )
\left ( - \langle 5 7 \rangle [8 6] \langle 2 3 1]
+ \langle 2 5 \rangle [ 6 1 ] \langle 7 3 8] - \langle 2 7 \rangle [ 8 1 ] \langle 5 4 6 ] \right ),
\\
& A_{+}^{Z/\gamma} = 4 F_V(s) \left ( 2 \sin^2 \theta_W \; Q_{u} D_{\gamma}(s) + (V_{u} - A_{u}) D_Z(s) \right )
\left ( - \langle 5 7 \rangle [8 6] [ 2 3 1 \rangle
+ [ 2 6] \langle 5 1 \rangle \langle 7 3 8] - [ 2 8] \langle 7 1 \rangle \langle 5 4 6 ] \right ),
\end{split}
\ee
where $\theta_W$ is the weak mixing angle, $Q_u = 2/3$ is the up quark electric chage, $V_{u} = 1/2-4/3 \sin^2 \theta_W$
and $A_u =1/2$ are vector and axial couplings of the up quark to $Z$-boson, $D_{\gamma } = 1/s$ and $D_Z = 1/(s-m_Z^2)$,
respectively. The $Z/\gamma$ part of the amplitude is proportional to the quark form-factor of the vector current
$F_V(s)$ whose perturbative expansion through NNLO is well-known.
Finally, we note that in addition to $u \bar u$ amplitude we also need the $d \bar d$ amplitude to compute the cross-section.
This amplitude can be easily obtained from the previous one by interchanging $W^+$ with $W^-$. This requires
changing $t \leftrightarrow u$ and $p_3^2 \leftrightarrow p_4^2$ in the form-factors $F_{1,2,...15}$ and
$p_3 \leftrightarrow p_4,\;\; p_5 \leftrightarrow p_7$ and $p_6 \leftrightarrow p_8$ in the amplitude
Eq.(\ref{eqampl}).