Abstract
Based on the method developed in Phan and Riemann (Phys Lett B 791:257, 2019), detailed analytic results for scalar one-loop two-, three-, four-point integrals in general d-dimension are presented in this paper. The calculations are considered all external kinematic configurations and internal mass assignments. Analytic formulas are expressed in terms of generalized hypergeometric series such as Gauss \(_2F_1\), Appell \(F_1\) and Lauricella \(F_S\) functions.
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1 Introduction
Scalar one-loop integrals in general d are important for several reasons. In general framework for computing two-loop and higher-loop corrections, higher-terms in the \(\epsilon \)-expansion (\(\epsilon =2-d/2\)) for one-loop integrals are necessary for building blocks. For example, they are used for building counterterms. Furthermore, in the evaluations for multi-loop Feynman integrals, we may combine several methods in [1, 2] to optimize the master integrals. As a result, the resulting integrals may include of one-loop functions in arbitrary space-time dimensions. Last but not least, scalar one-loop functions at \(d= 4+2n \pm 2\epsilon \) with \(n\in {\mathbb {N}}\) may be taken into account in the reduction for tensor one-loop Feynman integrals [3].
One-loop functions in space-time dimensions d have been performed in [4,5,6,7,8,9]. However, not all of the calculations have covered at general configurations of external momenta and internal masses. Recently, scalar one-loop three-point functions have been performed by applying multiple unitarity cuts for Feynman diagrams [10]. In Ref. [10], analytic results have been presented in terms of hypergeometric functions in special cases of external invariants and internal masses. In more general cases, the results have only presented \(\epsilon ^0\)-terms in space-time \(d=4-2\epsilon \). The algebraic structure of cut Feynman integrals, the diagrammatic coaction and its applications have proposed in [11,12,13]. However, the detailed analytic results for one-loop Feynman integrals have not been shown in the mentioned papers. More recently, detailed analytic results for one-loop three-point functions which are expressed in terms of Appell \(F_1\) hypergeometric functions have been reported in [15].
A recurrence relation in d for Feynman loop integrals has proposed by Tarasov [2, 14, 16]. By solving the differential equation in d, analytic results for scalar one-loop integrals up to four-points have been expressed in terms of generalized hypergeometric series such as Gauss \({}_2F_1\), Appell \(F_1\), Lauricella–Saran \(F_S\) functions. In [16], boundary terms have obtained by applying asymptotic theory of complex Laplace-type integrals. These terms are only valid in sub-domain of external momentum and mass configurations which the theory are applicable. Hence, the general solutions for one-loop integrals in arbitrary kinematics have not been found [16], as pointed out in [17, 18]. General solutions for this problem have been derived in [19] which proposed a different kind of recursion relation for one-loop integrals in comparison with [2, 14, 16]. In the scope of this paper, based on the method in Ref. [19], detailed analytic results for scalar one-loop two-, three- and four-point functions in general d-dimensions are presented. The calculations are considered all external kinematic configurations and internal mass assignments. Thus, we go beyond the material presented in [19].
The layout of the paper is as follows: In Sect. 2, we discuss briefly the method for evaluating one-loop integrals. We then apply this method for computing scalar one-loop two-, three- and four-point functions in Sects. 3, 4, and 5. Conclusions and plans for future work are presented in Sect. 6.
2 Methods
In this section, we describe briefly the method for evaluating scalar one-loop N-point Feynman integrals. Detailed description for this method can be found in Ref. [19]. A general recursion relation between scalar one-loop N-point and (N-1)-point Feynman integrals is shown in this section. From the representation, analytic formulas for scalar one-loop N-point functions can be constructed from basic integrals which are scalar one-point integrals. For illustrating, analytic expressions for scalar one-loop two-, three-, four-point functions are derived in detail in next sections.
The scalar one-loop N-point Feynman integrals are defined:
The inverse Feynman propagators are given:
where \(p_j\) (\(m_j\)) for \(j=1,2,\ldots , N\) are external momenta (internal masses) respectively. The momenta \(q_j\) are given: \(q_1 =p_1, q_2 =p_1+p_2, \ldots , q_j = \sum _{i=1}^{j}p_i\) and \(q_N=\sum _{j=1}^{N}p_j=0\) thanks to momentum conservation. They are inward as described in Fig. 1. The term \(i\rho \) is Feynman’s prescription and d is space-time dimension. Several cases of physical interests for d are \(d= 4+2n \pm 2\epsilon \) with \(n\in {\mathbb {N}}\). In the complex-mass scheme [20], the internal masses take the form of \(m_j^2 = m_{0j}^2-i m_{0j}\Gamma _{j}\) where \(\Gamma _{j}\geqslant 0\) are decay widths of unstable particles.
The Cayley and Gram determinants [16] related to one-loop Feynman N-point topologies are defined as follows:
with \(Y_{ij}=-(q_i-q_j)^2+m_i^2+m_j^2\).
In this report, analytic solutions for one-loop integrals are expressed in terms of generalized hypergeometric with arguments given by ratios of the above determinants. Hence, it is worth to introduce the following kinematic variables
The kinematics \(R_N\) play a role of the squared internal masses. In fact, when we shift \(m_j^2 \rightarrow m_j^2 - i\rho \), one verifies easily that \(R_N \rightarrow R_N- i\rho \) [16].
The recursion relation for \(J_N\) [19] is given (master equation):
for \(i,j=1,2,\ldots , N\) and \(\partial _k=\partial /\partial m_k^2\). Here the operator \(\mathbf{k}^-\) is defined as [19]
The relation (5) indicates that the integral \(J_N\) can be constructed by taking one-fold Mellin–Barnes (MB) integration over \(J_{N-1}\) in \(d+2s\). This representation has several advantages. First, analytic formulas for \(J_N\) can be derived from basic functions which are scalar one-loop one-point functions. Second, \(J_N\) is expressed as functions of kinematic variables such as \(m_j^2\) for \(j=1,2,\ldots , N\) and \(R_N\). As a consequence of this fact, analytic expressions for \(J_N\) reflect the symmetry as well as threshold behavior of the corresponding Feynman topologies. Two special cases of (5) are also mentioned as follows:
- 1.
\(Y_N \rightarrow 0 \) and \(G_{N-1} \ne 0\): In this case, \(R_N \rightarrow 0\) and we have [16] (deriving this equation for \(N=4\) is shown in the Appendix D)
$$\begin{aligned}&J_N(d;\{p_ip_j\}, \{m_i^2\}) = \frac{1}{d-N-1}\nonumber \\&\quad \times \sum \limits _{k=1}^N \left( \frac{\partial _k Y_N}{G_{N-1}} \right) \mathbf{k^{-} } J_N(d-2;\{p_ip_j\}, \{m_i^2\}). \end{aligned}$$(7) - 2.
\(G_{N-1} \rightarrow 0\) and \(Y_N \ne 0\): In this case, \(R_N \rightarrow \infty \). We close the integration contour in (5) to the right. Taking residue contributions from poles of \(\Gamma (\cdots -s)\). In the limit \(R_N \rightarrow \infty \), we find only the term with \(s=0\) is non-zero. The result then reads
$$\begin{aligned}&J_N(d;\{p_ip_j\}, \{m_i^2\}) \nonumber \\&\quad = -\frac{1}{2}\sum \limits _{k=1}^N \left( \frac{\partial _k Y_N }{Y_N} \right) \; \mathbf{k}^{-} J_N(d;\{p_ip_j\}, \{m_i^2\}). \end{aligned}$$(8)This equation is equivalent to (65) in [21] and (3) in [16].
We turn our attention to apply the method for evaluating scalar one-loop Feynman integrals. The detailed evaluations for scalar one-loop two-, three- and four-point functions are presented in next sections. As we pointed out in this section, the prescription \(i\rho \) always follows with \(R_N\) as \(R_N-i\rho \). In order to simplify the notation, we omit \(i\rho \) in \(R_N\) in the next calculations. This term puts back into the final results when it is necessary.
3 One-loop two-point functions
The master equation for \(J_2\) is obtained by setting \(N=2\) in (5). MB representation for \(J_2\) then reads
Note that we used the analytic formula for \(J_1\) in [22] with d shifted to \(d\rightarrow d+2s\). We write \(J_1\) in \(d+2s\) explicitly as follows:
In order to evaluate the MB integrals in (9), we close the integration contour to the right. The residue contributions to \(J_2\) at the sequence poles of \(\Gamma (-s)\) and \(\Gamma \left( \frac{2-d}{2}-s\right) \) are taken into account.
First, we calculate the residue at the poles of \(\Gamma (-s)\). In this case, \(s = m\) for \(m=0,1, \ldots , {\mathbb {N}}\). Subsequently, we can apply the reflect formula for gamma functions (103) in the Appendix B. In detail, it is implied that
With the help of (11), the MB representation in (9) is casted into the form of
Using (1.6.1.6) in Ref. [25], these MB integrals are expressed in terms of Gauss hypergeometric series as follows:
provided that \(\left| m_1^2/R_2\right| <1\), \(\left| m_2^2/R_2\right| <1\) and \({\mathcal {R}}\)e\((d-2)>0\). Using (1.3.15) in [25], we arrive at another representation for (13)
provided that \(\left| m_1^2/R_2\right| <1\), \(\left| m_2^2/R_2\right| <1\) and \({\mathcal {R}}\)e\((d-2)>0\).
We next consider the residue at the second sequence poles of \(\Gamma \left( \frac{2-d}{2} -s\right) \). In this case, \(s =\frac{2-d}{2}+m \) for \(m \in {\mathbb {N}}\). These contributions read
Noting that from (15) to (16), we have already applied the reflect formulas for gamma functions [see (103) in Appendix A for more detail]. Summing all the above contributions in Eqs. (14, 17), we finally get
provided that \(\left| m_1^2/R_2\right| <1\), \(\left| m_2^2/R_2\right| <1\) and \({\mathcal {R}}\)e\((d-2)>0\). Equation (18) is unchanged with exchanging \(m_1^2 \leftrightarrow m_2^2\). This reflects the symmetry of the scalar one-loop two-point Feynman diagrams. The result in (18) has shown in [18] and gives fully agreement with [16]. It is an important to remark that the solution for \(J_2\) in (18) is also valid when \(d\rightarrow d+2n\) for \(n\in {\mathbb {N}}\). We would like to stress that one can perform the analytic continuation for \(J_2\) in (18) to extend the kinematic regions for one-loop two-point functions. The analytic continuation formulas for Gauss hypergeometric functions are given from (109) to (112) in the Appendix B. As an example, using (111), the result reads
With (108), we arrive at
We are going to consider special cases for scalar one-loop two-point Feynman integrals.
3.1 \(G_1 \ne 0\) and \(R_2 = 0\)
If \(R_2 =0\), we verify that \(p^2 = (m_1-m_2)^2\), or \((m_1+m_2)^2\). Applying (7) for \(N=2\) the result reads
In the limit of \(m_1\rightarrow m_2 =m\) and \(p^2 =4m^2\), the result reads
3.2 \(G_1 = 0\)
Following (8) for \(N=2\), we arrive at
If \(m_1^2 =m_2^2\), one presents \(J_2\) as
3.3 \(R_2= m_1^2\) or \(R_2=m_2^2\)
For the case of \(R_2= m_1^2\) or \(R_2=m_2^2\), one relies on (20). If \(R_2 =m_1^2= m_2^2=m^2\), the result in (20) simplifies to
Under the condition \({\mathcal {R}}\)e\((d-4)>0\), when \(m^2 \rightarrow 0\), one then gets \(J_2=0\).
3.4 \(m_1^2=m_2^2=0\)
If one of \(m_i^2=0\), for \(i=1,2\), we rely on (18). In the case of \(m_1^2=m_2^2=0\), from (18) we get
provided that \({\mathcal {R}}\)e\(\left( d-2\right) >0\). We note that \(p^2\) means \(p^2 + i\rho \). Therefore, if \(p^2>0\) the term \(\left( -p^2/4\right) ^{\frac{d-4}{2}}\) is well-defined.
4 One-loop three-point functions
Setting \(N=3\) in (5), master equation for \(J_3\) reads
for \(i=1,2,3\). The term \(\mathbf{k}^- J_3(d+2s; \{p_ip_j\}, \{m_i^2\})\) becomes scalar one-loop two-point functions by shrinking an propagator k-th in the integrand of \(J_3\). In the next steps, we take the contour integrals in (27). In order to understand how to take the contour integrals, we chose the term with \(k=3\) in (27) for illustrating. This term is written explicitly as follows:
For taking the MB integrals in (29), one closes the integration contour to the right. The residue contributions at the poles of the \(\Gamma (\cdots -s)\) are taken into account.
First, the contributions of residua at the poles \(s = m\) with \(m\in {\mathbb {N}}\). In this case, one first applies the reflect formula (103) for gamma function:
Using this identity, the first MB integration reads
This MB integral is then expressed in terms of hypergeometric \(_2F_1\) as follows:
provided that \(\left| R_{12}/R_3\right| <1\) and \({\mathcal {R}}\)e\(\left( d-3\right) >0\). The second MB integral reads
provided that \(\left| m_{(1,2)}^2/R_3\right| \), \(\left| m_{(1,2)}^2/R_{12} \right| <1\) and \({\mathcal {R}}\)e\(\left( d-2\right) >0\).
In the next steps, the residue contributions at the second sequence poles \(s = \frac{4-d}{2}+ m\) for \(m\in {\mathbb {N}}\) are taken into account. The next MB integrations are considered as follows:
provided that \(\left| m_{(1,2)}^2/R_3\right| \), \(\left| m_{(1,2)}^2/R_{12}\right| <1\) and \(\left| R_{12}/R_3\right| <1\).
Summing all the above contributions, the final result for \(J_3\) is written as a compact form
where \(J_{123}^{(d)}\) is obtained from (32) and (33). It is given by
provided that \(\left| m_{i}^2/R_{ij} \right| <1\), \(\left| R_{ij}/R_{ijk} \right| <1\) for \(i,j,k=1,2,3\) and \({\mathcal {R}}\)e\(\left( d-2 \right) >0\). The latter condition always meets when \(d>2\). The kinematic variables \(R_{ijk}, R_{ij}\) and \(m_{i}\) for \(i,j,k =1,2,3\), etc., may not satisfy the former conditions. If the absolute value of the arguments of \(\,{}_2F_1\) and the Appell functions \(F_1\) in (39) are larger than one, we have to perform analytic continuations for these functions as in [25, 31]. The result for \(J_3\) has been shown in [18, 19]. The term \(J_{123}^{(d=4)}\) is obtained from (36, 37) or taking \(d\rightarrow 4\) of (39). This term is given
We emphasis that the solution (38) for \(J_3\) with (39) is equivalent to (74) in Ref. [16]. But the terms \(J^{(d=4)}_{123}, \ldots \) in our solution cover the condition (73) in Ref. [16]. Since the boundary term given in (74) of Ref. [16] was obtained by asymptotic theory of complex Laplace-type integrals. This term is only valid in a kinematic sub-domain in which the asymptotic theory of Laplace-type can be applied. The analytic continuation for the boundary term in [16] has not been discussed. We provide a complete analytic solution for \(J_3\) in comparison with (74) in Ref. [16]. We refer to our previous work [18] in which the numerical studies for this problem have discussed.
One-fold integral and all transformations for \(F_1\) can be found in Appendix B. Applying the relation (122) for \(F_1\) in Appendix C, we arrive at another representation for (39, 40):
provided that the absolute value of the arguments of \(\,{}_2F_1\) and the Appell functions \(F_1\) in this presentation are less than 1.
4.1 Massless internal lines
For the massless case, under the condition \({\mathcal {R}}\)e\((d-2)>0\), all terms related to Appell \(F_1\) functions in (38) vanish, the result then reads
In order to cross check with the result in [6], we write \(J_3\) as a function of \(p_1^2, p_2^2, p_3^2\) explicitly
Here \(\lambda (x,y,z) =x^2+y^2+z^2-2xy-2xz-2yz\) is the Källén function. We remark that \(p_i^2 \rightarrow p_i^2 + i\rho \) in this formula. With applying (1.3.3.5) in Ref. [25], one can present \(J_3\) as
provided that \(\left| \dfrac{\lambda (p_1^2, p_2^2, p_3^2)}{(p_2^2+p_3^2-p_1^2)^2} \right| <1\) and \({\mathcal {R}}\)e\((d-2)>0\). This equation is equivalent to (10) in [6]. We note that we can arrive to this result by inserting \(J_2\) at \(d+2s\) in (26) into (27) and taking the corresponding MB integrals.
4.2 \(R_{ij}=0\)
We consider the terms in \(J_3\) with \(R_{12} =0\) as an example. In this case, the terms \(J_{231}^{(d)}\) and \(J_{312}^{(d)}\) are given in the same form of (39) or (41). While the term \(J_{123}^{(d)}\) is obtained by performing analytic continuation the result in (41). In detail, one takes the limit of \(R_{12}\rightarrow 0\) in (41), we arrive at
Using (36) in Ref. [30], the term \(J_{123}^{(d)}\) simplifies to
provided that \({\mathcal {R}}\)e\(\left( d-2\right) >0\) and \(\left| m_i^2/R_3\right| <1\) for \(i=1,2\). Taking \(d\rightarrow 4\), we have
4.3 \(R_{ij}=m_{i(j)}^2\) for \(i,j=1,2,3\)
Next we consider \(R_{12} = m_1^2\) as an example. In this case, the terms \(J_{231}^{(d)}, J_{312}^{d}\) are given by (39). Beside that, one verifies
As a result, we obtain
provided that the amplitude of arguments of hypergeometric functions appearing in this formula are less that 1. For \(R_{12} = m_1^2=m_2^2\), the function \(F_1\) in (50) is equal 1. The result reads
4.4 \(R_3=m_k^2\) for \(k=1,2,3\)
As an example, consider the terms of \(J_3\) in (38) with \(R_3 =m_1^2\). One verifies that
As a result, \(J_3\) is casted into the form of
with \(J_{231}^{(d)}\) taking the same form of (39) or (41). We take (41) as example for \(J_{231}^{(d)}\). In detail, it takes
Taking \(d\rightarrow 4\), the result reads
4.5 \(G_2\ne 0 \) and \(R_3 = 0\)
By setting \(N=3\) in (7), one obtains
The resulting of \(\mathbf{k^{-} } J_3(d-2; \{p_i^2\}, \{m_i^2\})\) is \(J_2\) in (18) or in (20) with \(d \rightarrow d-2\). As an example, we take \(J_2\) in (20) at \(d-2\). The result reads
We can confirm (57) again by using analytic continuation result of \(J_3\) in (41). In fact, when \(R_3 \rightarrow 0\) the Eq. (41) becomes
Plugging (61) into (38), we arrive at (58).
4.6 \(R_3 = R_{ij}\) for \(i,j =1,2,3\)
We consider the case of \(R_3 =R_{12}\) as an example. In this case, one verifies that
As a result, \(J_3\) is
Here \(J_{231}^{(d)}\) takes the same form as (39) or (41). We take (41) as example for \(J_{231}^{(d)}\). In detail, it takes
4.7 \(G_2 = 0\)
Setting \(N=3\) in (8), the result reads
This equation is equivalent to (46) in Ref. [21]. The term \(\mathbf{k^{-} } J_3(d; \{p_i^2\}, \{m_i^2\})\) corresponds to \(J_2\) in (20) as an example. We obtain
4.8 \(G_{1(ij)} = 0\) for \(i,j=1,2,3\)
\(G_{1(ij)}\) are the Gram determinants of two-point functions which are obtained by shrinking a propagator \(k\ne i,j\) in the three-point ones. Taking \(G_{1(12)}=0\) as an example, the term \(J_{123}^{(d)}\) is evaluated as follows. We put \(J_2\) in (23) into (27). Taking the corresponding MB integrations, the results read as form of (38) with
It is valid under the conditions that the arguments of the hypergeometric fuctions appearing in this formula are less that 1 and \({\mathcal {R}}\)e\((d-2)>0\).
4.9 Cross check with other papers
We consider \(p_2^2=p_3^2=0;p_1^2\ne 0, m_1^2=m_3^2=0\) and \(m_2^2\ne 0\) as an example [10, 15]. We confirm that
The \(J_3\) in (27) becomes
We note that the first term in curly bracket of (71) is \(J_2\) in the case of (26) with d shifted to \(d+2s\). While the second term in curly bracket of (71) is corresponding to \(J_2\) in (23) at \(d+2s\) (and with a massless internal line). In the following we perform the contour integration of (71) starting with the second contour integral:
By closing the integration contour to the right, the residue contributions at the poles of \(\Gamma (-s)\) and \(\Gamma \left( \frac{4-d}{2} - s\right) \) are calculated. For the first sequence poles, the result reads
For the second sequence poles, we arrive at
Second type of MB integral is considered
For the first sequence poles of \(\Gamma (-s)\), the result is
Applying transforms for Gauss hypergeometric function which are (see (1.8.10) in [25] for the first relations and page 49, [26] for the later case)
one obtains
Taking into account the residue at the poles \(\Gamma \left( \frac{4-d}{2}-s \right) \), we get
Using the relation (see Eq. (3.1.7) in [27])
one gets
Combining all the terms, \(J_3\) reads
provided that \(\left| p_1^2/m_2^2\right| <1\) and \({\mathcal {R}}\)e\((d-2)>0\). It agrees with Eq. (B2) in [10] and (22) in [15].
5 One-loop four-point functions
The master equation for \(J_4\) is obtained from (5) with \(N=4\),
We substitute the analytic solution for \(J_3(d+2s; \{p_i^2\},\{m_i^2\}) \) in (38) into (85) and take the contour integrals in (85). With the help of MB integrations in (131, 134) in Appendix C, a compact expression for \(J_4\) can be derived and expressed as follows:
with
where \(J_{123}^{(d=4)}, \ldots \) are given by (40). It is important that this representation is valid under the conditions that \({\mathcal {R}}\)e\(\left( d-3\right) >0\) and the absolute values of arguments of hypergeometric functions are smaller than one. If the absolute value of these arguments are larger than one, we have to perform analytic continuations for the Gauss hypergeometric and Appell \(F_1\) functions, cf. [25, 31]. Further, the Saran function \(F_S\) may be expressed by a Mellin–Barnes representation, or Euler integrals in this case. The result for \(J_4\) has been shown in [19]. There are two important points we would like to emphasize in this paper as follows. (1) Ref. [16] have not shown conditions for the boundary term in (100). (2) \(J_4\) is constructed from \(J_3\) for arbitrary kinematics. However, the boundary term for \(J_3\) for general kinematics have not been provided in [16], as mentioned in the previous section and in [18]. Subsequently, the first term in (99) of [16] is only valid in special kinematic regions. Therefore, the solution in (98) of Ref. [16] may not be considered as a complete solution for \(J_4\). In contrast to [16], we provide a complete solution for \(J_4\) in this article.
5.1 Massless internal lines
We are going to take \(m_i^2 \rightarrow 0\) for \(i=1,2,3,4\). The terms related to \(F_S\) vanish. Therefore, in the massless case the result reads
provided that \({\mathcal {R}}\)e\(\Big (d-3\Big )>0\) and that the absolute values of arguments of the hypergeometric functions are smaller than one. Taking \(d\rightarrow 4\), we have
This is a new result for \(J_4\) in the massless case at general d. We are going to consider the special cases for \(J_4\) in the following subsections.
5.2 \(R_4 = 0\)
From (7), we set \(N=4\) and get
The term \(\mathbf{k^{-} } J_4(d-2;\{p_i^2, s, t\}, \{m_i^2\})\) is given by \(J_3\) in (38) with \(d\rightarrow d-2\).
5.3 \(R_4 = R_{ijk}\) for \(i,j,k =1,2,3, 4\)
As an example, we consider the case \(R_4 = R_{123}\). In this case, we verify that
As a result, the terms \(J_{1234}^{(d)}\) vanish, other terms in (86) are of the same form in (87).
5.4 \(R_4 = R_{ij}\) for \(i,j = 1,2,3,4\)
For example, the terms of \(J_4\) in (86) meet the condition \(R_4 = R_{12}\). Because that \(R_2\) depends only the internal masses \(m_1^2, m_2^2\), one has
As a matter of this fact, only two terms (2, 3, 4, 1) and (3, 4, 1, 2) in (86) contribute to \(J_4\).
5.5 \(R_4 = m_k^2\) for \(k=1,2,3,4\)
For example, one considers the terms of \(J_4\) in (86) having \(R_4 =m_1^2\). One verifies that
As a result, only the term (2, 3, 4, 1) in (86) contributes to \(J_4\).
5.6 \(R_{ijk} = 0\) for \(i,j,k=1,2,3,4\)
We assume that \(J_4\) in (86) contains \(R_{123} =0\) as an example. The term (1, 2, 3, 4) in (87) with \(R_{123}=0\) is evaluated by applying the same previous procedure. The result reads
Where \(F^{1; 2;1}_{1;1;0}\) is Kamp\(\acute{\text {e}}\) de F\(\acute{\text {e}}\)riet [33] (see Appendix B for more detail). We also refer to [34] which analytic continuations for a class of the Kamp\(\acute{\text {e}}\) de F\(\acute{\text {e}}\)riet functions have been studied. This representation is valid if the amplitude of arguments of these hypergeometric functions are less than 1 and \({\mathcal {R}}\)e\((d-4)>0\). In the massless case, one has
5.7 \(R_{ijk} = R_{ij}\) for \(i,j, k=1,2,3,4\)
We examine the terms of \(J_4\) in (86) having \(R_{123} = R_2\). We check that
As a result, the terms (2, 3, 4, 1), (3, 4, 1, 2) and (4, 1, 2, 3) of \(J_4\) get the same formula for \(J_{1234}^{(d)}\) provided in (87). The terms (1, 2, 3) of \(J_{1234}^{(d)}\) in (87) are vanished.
5.8 \(R_{ijk} = m_{i,(j,k)}^2\) for \(i,j,k = 1,2,3,4\)
Assuming the terms of \(J_4\) in (86) with \(R_{123} = m_1^2\), we check that
The terms \(J_{2341}^{(d)}, J_{3412}^{(d)}\) and \(J_{4123}^{(d)}\) get the same formula as (87). The term (1, 2, 3) and (3, 1, 2) of \(J_{1234}^{(d)}\) in (87) vanish, only the (2, 3, 1)-term contribute to \(J_{1234}^{(d)}\).
5.9 \(R_{ij} = 0\) for \(i,j=1,2,3,4\)
We consider that \(J_4\) in (86) with \(R_{12} =0\). In this case, the terms \(J_{2341}^{(d)}\) and \(J_{3412}^{(d)}\) are unchanged. The terms \(J_{1234}^{(d)}\) and \(J_{4123}^{(d)}\) with \(R_{12} = 0\) are evaluated again by applying same previous procedure. Taking \(J_{1234}^{(d)}\) as an example. The result reads
This representation is valid if the amplitude of arguments of these hypergeometric functions are less than 1 and \({\mathcal {R}}\)e\((d-3)>0\).
5.10 \(R_{ij} = m_i^2\) or \(m_j^2\) for \(i,j=1,2,3,4\)
Taking \(R_{12} = m_1^2\) as an example, one confirms that
As a result, the terms (1, 2, 3) of \(J_{1234}^{(d)}\) that are multiplied by \(\frac{\partial R_2}{\partial m_2^2}\) vanish. Other terms of \(J_{1234}^{(d)}\) are given by (87). The terms \(J_{2341}^{(d)}, J_{3412}^{(d)}\) and \(J_{4123}^{(d)}\) of \(J_4\) are given by (87).
5.11 \(G_3 = 0\)
In this case, one has
This equation is equivalent with (65) in Ref. [21]. The term \(\mathbf{k}^{-} J_4(d;\{p_i^2, s, t\}, \{m_i^2\})\) is given by \(J_3\) in (38).
5.12 \(G_{2(ijk)} = 0\) for \(i,j,l =1,2,3,4\)
In the same notation, \(G_{2(ijk)}\) are the Gram determinants of \(J_3\) that are obtained by shrinking kth propagator in \(J_4\). We take \(|G_{2(123)}| = 0\) as an example. By using (66) for \(J_{123}^{(d)}\), we then evaluate \(J_{1234}^{(d)}\) again, the result is
This representation is valid if the amplitude of arguments of these hypergeometric functions are less than 1 and \({\mathcal {R}}\)e\((d-3)>0\).
5.13 \(G_{1(ij)}= 0\) for \(i,j=1,2,3,4\)
One assumes that the term \(J_{1234}^{(d)}\) has \(G_{1(12)}=0\). Recalculating this term, the result reads in term of Gauss and Appell \(F_3\) functions
where the terms \(J_{123}^{(d=4)}\) and \(J_{231}^{(d=4)}\) are given in (40). This representation is valid if the amplitude of arguments of these hypergeometric functions are less than 1 and \({\mathcal {R}}\)e\((d-3)>0\). The Appell \(F_3\) functions are described in detail in Appendix B (see (123) in more detail).
For future prospect of this work, a package which provides a general \(\epsilon \)-expansion and numerical evaluations for one-loop functions at general d is planned. To achieve this purpose, many related works are worth mentioning in this paragraph. First, automatized analytic continuation of Mellin–Barnes integrals have been presented in [38]. The construction of Mellin–Barnes representations for Feynman integrals has been performed in [39, 40]. Recent development for treating numerically Mellin–Barnes integrals in physical regions has been proposed in [41,42,43]. The hypergeometric functions in this work can be expressed as the multi-fold MB integrals and they may be evaluated numerically by following the above works. Furthermore, the \(\epsilon \)-expansion of the hypergeometric functions appearing in our analytic results may be also performed by using the packages Sigma, EvaluateMultiSums and Harmonic Sums [44,45,46,47,48,49,50]. Numerical \(\epsilon \)-expansion of hypergeometric functions may be done by using NumEXP [51]. Besides that, analytic \(\epsilon \)-expansion for the hypergeometric functions has been carried out in [52,53,54,55,56,57,58,59]. Differential reduction of generalized hypergeometric functions has been also reported in [60,61,62,63].
In the context of dimensional recurrence relations, the tensor reductions for one-loop up to five-point functions have been worked out in [64] and for higher-point functions have been developed in [65]. In practice, one encounters integrals with denominator powers higher than one and their reduction needs to be considered, see e.g. [3] for the scalar case. IBP reduction can be combined with dimensional recurrence relations to reduce them to master integrals of higher space-time dimensions.
6 Conclusions
In this article, we have been presented the analytic results for scalar one-loop two-, three- and four-point functions in detail. The results have been expressed in terms of Gauss \(_2F_1\), Appell \(F_1\) and \(F_S\) hypergeometric functions. All cases of external momentum and internal mass assignments have considered in detail in this work. The higher-terms in the \(\epsilon \)-expansion for one-loop integrals can be performed directly from analytic expressions in this work. These terms are necessary building blocks in computing two-loop and higher-loop corrections. Moreover, one-loop functions in arbitrary d in this work may be taken account in the evaluations for higher-loop Feynman integrals. The one-loop functions with \(d\geqslant 4\) can also used in the reduction for tensor one-loop Feynman integrals. For future works, a package for numerical evaluations for one-loop integrals at general d and general \(\epsilon \)-expansion for these integrals is planned. Additionally, the method can extend to evaluate two- and higher-loop Feynman integrals.
Data Availability Statement
This manuscript has no associated data or the data will not be deposited. [Authors’ comment: This is theoretical work, we present analytic results in this work. Therefore, we have not used any data.]
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Acknowledgements
This research is funded by Vietnam National Foundation for Science and Technology Development (NAFOSTED) under the Grant number 103.01-2019.346. The author would like to thank T. Riemann for helpful discussions and comments.
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Appendices
Appendix A: Useful formulas
In this appendix, we show some useful formulas used in this paper. We applied the reflect formula for gamma functions [24]:
provided that \(z\in {\mathbb {C}}\) and \(n \in {\mathbb {N}}\). We have mentioned the duplication formula for gamma functions [24]:
provided that \(z\in {\mathbb {C}}\).
Appendix B: Generalized hypergeometric series
Generalized hypergeometric functions are presented in this appendix.
1.1 Gauss hypergeometric series
Gauss hypergeometric series are given [25]:
provided that \(|z|<1\). The pochhammer symbol \((a)_n =\Gamma (a+n)/\Gamma (a)\) is used.
The integral representation for Gauss hypergeometric functions [25] reads
provided that \(|z|<1\) and Re\((c)>\)Re\((b)>0\).
Basic linear transformation formulas for Gauss \(_2F_1\) hypergeometric functions which are collected from Refs. [25, 28], are listed as follows:
1.2 Appell series
Appell \(F_1\) series are defined [25, 32]:
provided that \(|x|<1\) and \(|y|<1\).
The single integral representation for \(F_1\) reads [25]
provided that Re(c) > Re\((a)>0\) and \(|x|<1\), \(|y|<1\).
We collect all transformations for Appell \(F_1\) functions from Refs. [25, 30]. The first relation for \(F_1\) is mentioned,
If \(b'=0\), we arrive at the well-known Pfaff–Kummer transformation for the \(_2F_1\). Further, we have
Furthermore, if \(c=b+b'\), one then obtains
Similarly,
and
A further relation for \(F_1\) is given
1.3 Appell \(F_3\) series
Appell \(F_3\) series are written [25, 32]
provided that Re(c) > Re\((a)>0\) and \(|x|<1\), \(|y|<1\).
1.4 J. Kampé de Fériet series
In addition, J. Kampé de Fériet series [32, 33] with two variables are shown:
1.5 Lauricella–Saran function \(F_S\)
Lauricella–Saran function \(F_S\) is defined in terms of a triple hypergeometric series [35]
The integral representation of \(F_S\) is defined [35]:
provided that Re\((\gamma _1-\alpha _1-\alpha _2)>0\), Re\((\alpha _1)>0\), Re\((\alpha _2)>0\), and \(|x|,\;|y|,\; |z|<1\).
Appendix C: The contour integrations
1.1 Type 1
Mellin–Barnes relation [29] is given:
provided that \(|\mathrm {Arg}(z)|<\pi \). The integration contour is chosen in such a way that the poles of \(\Gamma (- s)\) and \(\Gamma (\lambda + s)\) are well-separated.
1.2 Type 2
The Barnes-type integral for Gauss hypergeometric [25] reads
provided that \(|\mathrm {Arg}(-z)|<\pi \) and \(|z|<1\).
1.3 Type 3
The next Barnes-type integral applied in this paper is
with \(|\mathrm {Arg}(-x)|<\pi \), \(|\mathrm {Arg}(-y)|<\pi \) and \(|x|<1\) and \(|y|<1\). Under these conditions, one could close the contour of integration to the right. Subsequently, we have to take into account the residua of the sequence poles of \(\Gamma (-s)\). The result is expressed as the summation of Gauss hypergeometric function. The summation is then identical as a series of Appell \(F_1\) functions [30] as follows:
1.4 Type 4
Furthermore, in this paper we evaluate the following integral:
We close the contour of integration to the right. By taking into account the residua of the sequence poles of \(\Gamma (-s)\), the result reads
One applies the relation (115) for Appell \(F_1\) functions in (132) as follows
Equation (132) is then presented as a series of Lauricella functions \(F_S\) [35]
provided that \(|x|<1\), \(\left| \frac{y}{y-1}\right| <1\) and \(\left| \frac{z}{z-1}\right| <1\).
The last formula is mentioned in this paper relates to a transformation of Mellin–Barnes integrals (see page 156 of Ref. [28] , item 14.53 in page 290 of [29]) which is
provided that \(|\mathrm {Arg}(-z)|<2\pi \).
Appendix D: Master equation for \(J_N\)
General relation for \(J_N\) have been proved in [18]. In this appendix, we consider \(N=4\) as an example. Performing Feynman parameterization for \(J_4\), one arrives at
Above coefficients \(a, b, c,\ldots ,j\) are shown
We used the following notation
The integrand of \(J_4\) is
These matrices are given
The \(\Lambda _4(x_1,x_2, x_3)\) reads
The vector \(\overrightarrow{y}\) is defined as: \(\overrightarrow{y}= (y_1, y_2, y_3) = - {\mathcal {G}}_3^{-1} H^T\). We write explicitly \(y_1, y_2, y_3\) and \(y_4\) as follows:
First, we consider \(G_3 \ne 0\) and \(R_4 \ne 0\). In this case, Mellin–Barnes relation is applied to decompose \(J_4\)’s integrand as follows:
provided that \(\left| \mathrm {Arg}\left( \frac{\Lambda _4(x_1,x_2,x_3)}{R_4-i\rho } \right) \right| <\pi \). With the help of the Mellin–Barnes relation, this brings the Feynman parameters integration to the simpler form:
In order to carry out this integral, we consider the following differential operator (see theorem of Bernshtein [36], or [37])
It is easy to check that
As a matter of this fact, we can rewrite Feynman parameter integral as
The last term in this equation is proportional to \({\mathcal {F}}_4\). It is then combined with \({\mathcal {F}}_4 \) on the left side of Eq. (148). As a result, Eq. (148) is then casted into the form:
Taking over a Feynman parameter integration in Eq. (149), the result reads
The coefficients \(A_i, B_i, C_i, D_i, E_i, F_i \) are given in Table 1.
We then write the coefficients in terms of external momenta and internal masses which are given in Table 2.
Now we can write the Mellin–Barnes integral for \(J_4\) as
By applying the relation (135), one gets
This equation is then written as follows:
In the case of \(R_4 = 0\), there is no Mellin–Barnes integral for \(J_4\). We only apply the ring operator \({\hat{O}}_4\), the result arrives
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Phan, K.H. Scalar 1-loop Feynman integrals as meromorphic functions in space-time dimension d, II: special kinematics. Eur. Phys. J. C 80, 414 (2020). https://doi.org/10.1140/epjc/s10052-020-7989-4
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DOI: https://doi.org/10.1140/epjc/s10052-020-7989-4