Gravitational lens without asymptotic flatness: Its application to the Weyl gravity

Keita Takizawa, Toshiaki Ono, and Hideki AsadaGraduate School of Science and Technology, Hirosaki University, Aomori 036-8561, Japan

(Dated: September 14, 2020)

We discuss, without assuming asymptotic flatness, a gravitational lens for an observer and sourcethat are within a finite distance from a lens object. The proposed lens equation is consistent withthe deflection angle of light that is defined for the nonasymptotic observer and source by Takizawaet al. [Phys. Rev. D 101, 104032 (2020)] based on the Gauss-Bonnet theorem with using the opticalmetric. This lens equation, though it is shown to be equivalent to the Bozza lens equation [Phys.Rev. D 78, 103005 (2008)], is linear in the deflection angle. Therefore, the proposed equationis more convenient for the purpose of doing an iterative analysis. As an explicit example of anasymptotically nonflat spacetime, we consider a static and spherically symmetric solution in Weylconformal gravity, especially a case that γ parameter in the Weyl gravity model is of the orderof the inverse of the present Hubble radius. For this case, we examine iterative solutions for thefinite-distance lens equation up to the third order. The effect of the Weyl gravity on the lensedimage position begins at the third order and it is linear in the impact parameter of light. Thedeviation of the lensed image position from the general relativistic one is ∼ 10−2 microarcsecondsfor the lens and source with a separation angle of ∼ 1 arcminute, where we consider a cluster ofgalaxies with 1014M at ∼ 1 Gpc for instance. The deviation becomes ∼ 10−1 microarcseconds,even if the separation angle is ∼ 10 arcminutes. Therefore, effects of the Weyl gravity model arenegligible in current and near-future observations of gravitational lensing. On the other hand, thegeneral relativistic corrections at the third order ∼ 0.1 milliarcseconds can be relevant with VLBIobservations.

PACS numbers: 04.40.-b, 95.30.Sf, 98.62.Sb

I. INTRODUCTION

The gravitational deflection of light has been an impor-tant tool in gravitational physics, since it was measuredby Eddington and his collaborators [1]. Gravitationallens has been one of the key subjects in the modern as-tronomy and cosmology, though Einstein thought thatthe phenomenon of a star acting as a gravitational lenswas unobservable [2]. In particular, the Event HorizonTelescope (EHT) team has recently succeeded a directimaging of the immediate vicinity of the central blackhole candidate of M87 galaxy [3].

The formulation of the gravitational lens and its appli-cations are usually based on the gravitational lens equa-tion. The conventional lens equation uses the deflectionangle of light that is defined for the asymptotic receiver(denoted by R) and source (denoted by S), where theobserver is referred to the receiver in order to avoid aconfusion in notations between r0 (the closest approachof light) and rO by using rR.

Gibbons and Werner proposed an alternative way ofdefining the asymptotic deflection angle of light [4], wherethe receiver and source of light are assumed to be in anasymptotically Minkowskian region. The Gauss-Bonnettheorem [5] with using the optical metric plays a crucialrole in their geometrical definition of the deflection an-gle. Their method has been vastly applied to a lot ofspacetime models especially by Jusufi and his collabo-rators e.g. [6–8], and has been extended to study thegravitational deflection of light in a plasma medium e.g.[9, 10].

Ishihara et al. extended the idea of Gibbons and

Werner to study effects of finite distance on the gravi-tational deflection of light, where the receiver and sourceare within a finite distance from a lens object [11, 12].Their formulation has been extended to stationary andaxisymmetric spacetimes such as Kerr solution [13], a ro-tating wormhole [14] and a rotating global monopole withan angle deficit [15]. Their definition of the deflection an-gle is still limited within asymptotically flat spacetimes.See Reference [16] for a review on this subject.

Without assuming asymptotic flatness, Takizawa et al.proposed a definition of the gravitational deflection oflight for the receiver and source that are within a finitedistance from a lens object [17]. In their definition basedon the Gauss-Bonnet theorem, the radial interval is ex-actly the same as that for the light ray from the sourceto the receiver. As a result, this definition can be ap-plied not only to an asymptotically flat black hole butalso to an asymptotically nonflat black hole such as theKottler (Schwarzschild-de Sitter) solution in general rel-ativity and a static and spherically symmetric vacuumsolution in Weyl conformal gravity.

The deflection angle of light is not always observable.As mentioned above, the gravitational lensing observ-ables are discussed by using the gravitational lens equa-tion. How can the deflection angle of light for nonasymp-totic receiver and source be incorporated into the gravi-tational lens equation? The main purpose of this paperis to discuss a gravitational lens equation valid for thedeflection angle of light that is defined by Takizawa etal. [17], without assuming asymptotic flatness, for anobserver and source within a finite distance from a lensobject.

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FIG. 1. ΩR and ΩS . ΩR is a trilateral specified by the pointsR, P0 and PR. ΩS is that specified by the points S, P0 andPS .

This paper is organized as follows. In Section II, thelens equation with finite-distance effects is reexamined.In Section III, we discuss iterative solutions for the finite-distance lens equation in the small angle approxima-tion. Section IV discusses the lensed image positions ina static, spherically symmetric vacuum solution in Weylconformal gravity. In Section V, we examine whether ef-fects of Weyl conformal gravity on the gravitational lenscan be tested by present and near-future astronomicalobservations. Section VI is devoted to the conclusion.Throughout this paper, we use the unit of G = c = 1.

II. LENS EQUATION IN A FINITE-DISTANCESITUATION

A. Effect of finite distances on the lightpropagation

We follow References [11, 17] to consider a static andspherically symmetric spacetime. The metric reads

ds2 = gµνdxµdxν

= −A(r)dt2 +B(r)dr2 + C(r)dΩ2, (1)

where dΩ2 ≡ dθ2 + sin2 θdφ2 and φ is the azimuthal an-gle respecting the rotational symmetry. If we chooseC(r) = r2, then, r denotes the circumference radius.Henceforth, we choose the photon orbital plane as theequatorial plane without the loss of generality, becausethe spacetime is spherically symmetric.

In order to avoid requiring the asymptotic flatness of aspacetime, Takizawa et al. proposed an integral form ofthe definition for the deflection angle of light (denoted asαK) for an observer and source that are within a finite

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FIG. 2. The light ray and radial directions. The anglebetween the light ray and the radial direction at the receiveris ΨR and that at the source is ΨS . The coordinate anglebetween the receiver and the source is φRS = φR − φS .

distance from a lens object [17]. αK is defined as

αK ≡∫∫

ΩR+ΩS

KdS +

∫ PS

PR

κgd`+ φRS . (2)

The right-hand side of this equation contains the radialcoordinate r ∈ [r0, rR] or [r0, rS ], where r0 means theclosest approach of light. Indeed, this radial interval isexactly the same as that for the light ray from the sourceto the receiver. See Figure 1.

Without assuming the asymptotic flatness, theyproved that their definition agrees with another form ofthe deflection angle by Ishihara et al. [11] which assumedthe asymptotic flatness. Ishihara et al. [11] defined thedeflection angle of light as

αI ≡ ΨR −ΨS + φRS , (3)

where ΨS and ΨR are the angles between the radial di-rection and the light ray at the source position and atthe receiver position, respectively, and φRS is a coordi-nate angle between the receiver and source. See Figure2 for these angles.

It was shown that

αI = αK , (4)

holds in general for a static and spherically symmetricspacetime, especially even for an asymptotically nonflatcase [17].

B. Finite-distance expressions for the deflectionangle of light

We introduce the lens plane and the source one toexamine the gravitational lens equation. See Figure 3

3

LR

Sbb

DL DLSDS

ΨS

ΨR

β

Source PlaneLens Plane

θU

QαG

FIG. 3. Geometrical gravitational lensing setup.

for the gravitational lensing configuration in this paper,where the thin lens approximation is not used. The (redin color) solid curve in this figure shows the light ray fromthe source to the receiver. The angles ΨR and ΨS appearin Eq. (3). The tangents at the receiver and the sourceare denoted by the dotted lines in this figure. Thesetangent lines intersect at the point Q. Note that the in-tersection point Q is not necessarily in the lens plane.In the conventional formulation with the thin lens ap-proximation for the asymptotic receiver and source, theintersection point is often assumed implicitly to be on thelens plane. The assumption that the intersection pointis in the lens plane needs a symmetric configuration inwhich the receiver and source are equidistant from thelens. This additional assumption is made also in Virb-hadra and Ellis for their formulation of the almost exactlens equation, though this formulation is valid not onlyfor the weak deflection but also for the strong deflection[18]. See e.g. Figure 1 and the paragraph including Eqs.(1)-(3) in Reference [18]. From the aspect of the triangu-lar inequality, Dabrowski and Schunck realized difficultiesof using Virbhadra and Ellis lens equation and derived analternative lens equation [19]. However, Dabrowski andSchunck lens equation still relies upon the additional as-sumption that the intersection point Q lies on the lensplane. See Eq. (23) and Appendix in Reference [19]. Theadditional assumption of the intersection point lying onthe lens plane was argued also by Bozza [20].

DL, DS and DLS denote the angular diameter dis-tances from the receiver to the lens, from the receiver tothe source and from the lens to the source, respectively.The angular direction of the lensed image with respect tothe lens direction is denoted by θ and that of the intrinsicsource position is denoted by β. These angles θ and βare defined at the receiver point. θ equals to ΨR. Seealso Figure 3.

We consider a quadrilateral LRQS in Figure 3. Figure4 focuses on the quadrilateral LRQS. In this geometrical

L

S

ΨS

ΨR

θ

QαG

R

φRS

FIG. 4. Quadrilateral LRQS in the geometrical gravita-tional lensing configuration. This is corresponding to Figure3. The directional difference between the receiver and sourceis assumed to be φRS at the point L. The (red in color) solidcurve denotes the light ray from the source to the receiver.The angle between the two tangent lines in this figure is in-terpreted as the deflection angle of light. The deflection angleis denoted as αG.

configuration of the gravitational lensing, we define thedeflection angle of light αG as the angle at the point Qbetween these tangent lines. In the gravitational lensinginterpretation, the inner angle at the lens in LRQS isassumed to be φRS . For the quadrilateral, we obtain

θ + φRS + (π −ΨS) + (π − αG) = 2π, (5)

where we follow the gravitational lensing interpretationto assume that the sum of the inner angles in any convexquadrilateral is 2π. By using Eq. (5), we define αG as

αG ≡ θ −ΨS + φRS . (6)

From Eqs. (3) and (6), we find

αI = αG, (7)

where we use ΨR = θ. Therefore, αI defined by Eq. (3)can be safely interpreted as the deflection angle of light.

From Eqs. (4) and (7), we obtain the equivalence of thethree definitions of the deflection angle of light, namely

αG = αI = αK . (8)

In the following, we use αG to study the gravitationallens equation, because αG is written in terms of θ thatplays a crucial role in the gravitational lens equation.

Before going to detailed calculations of αG, we brieflymention another finite-distance expression of the deflec-tion angle (denoted as αRM ) computed by Richter andMatzner for the PPN metric [21]. The equivalence be-tween αI(= αG) and αRM was noticed by Crisnejo et al.[22]. See Figure 5 for the lensing setup in αRM . Figures

4

L

S

ΨS

θ

QαRM

R

φRS

η

θ+η

FIG. 5. Gravitational lensing setup in Richter and Matznermethod for the PPN metric [21]. As a reference, they assumea (green in color) dashed line from the receiver. Here, the(green in color) dashed line is supposed to be parallel to the(blue in color) dashed line that is tangent to the light rayat the source. The deflection angle can be defined as αRM ≡θ+η, though the sign convention for η in this figure is oppositeto that by Richter and Matzner

4 and 5 show αG = αRM . It is worthwhile to point outthat the definition of αRM needs a comparison betweenthe two parallel lines (in Figure 5) and hence αRM israther limited compared with αG.

C. Effect of finite distances on the gravitationallens equation

αG is a relation among angles, in which any distancedoes not explicitly appear. Therefore, we shall studysome relations between angles and distances. The lightray (red solid curve in Figure 3) is specified by the impactparameter of light (denoted as b). At the point R, this isdescribed by

b = LR sin θ. (9)

At the point S, it is expressed as

b = LS sin(π −ΨS). (10)

The impact parameter b is common to Eqs. (9) and(10), so that b can be eliminated as

LS sin(π −ΨS) = LR sin θ. (11)

This is solved for ΨS as

ΨS = π − arcsin

(LR

LSsin θ

). (12)

SU= RS sin β= DS tan β

LS

B

FIG. 6. Geometrical meaning of B. The angle B is definedby Eq. (18).

We consider the triangles RSU and LSU in Figure 3.The length SU is written in two ways as

SU = LS sin(π − φRS), (13)

SU = RS sinβ. (14)

By eliminating SU from these equations,

LS sin(π − φRS) = RS sinβ. (15)

Hence, we obtain

φRS = π − arcsin

(RS

LSsinβ

). (16)

Substituting Eqs. (12) and (31) into Eq. (6), we obtain

αG − θ = arcsin

(LR

LSsin θ

)− arcsin

(RS

LSsinβ

).

(17)

LR = DL is a constant in the gravitational lensing for-mulation. On the other hand, LS and RS are dependenton the source position described by the parameter β. Wethus rewrite them in terms of the angular diameter dis-tances DL, DS and DLS .

Let B denote the second term in the right-hand sideof Eq. (17). Namely, it is defined by

B ≡ arcsin

(RS

LSsinβ

). (18)

The term B means an angle in a triangle by Figure 6.Note that this triangle does not appear in the lensingconfiguration by Figure 3 as it is. The length of the basefor this triangle in Figure 6 becomes√

(LS)2 − (RS sinβ)2 = DLS . (19)

5

Here, we used

SU = RS sinβ

= DS tanβ, (20)

and

(LS)2 − (SU)2 = (DLS)2. (21)

The first relation is obtained by using Figure 3 and thesecond one can be derived from the triangle LSU.

Eq. (19) is used for the triangle in Figure 6. B is thusrewritten in terms of the angular distances as

B = arctan

(DS

DLStanβ

). (22)

In the similar manner, we use Eqs. (20) and (21) toobtain

LS =√

(DLS)2 + (SU)2

=√

(DLS)2 + (DS)2 tan2 β. (23)

Eqs. (22) and (23) are substituted into the second andfirst terms in the right-hand side of Eq. (17), respectively.We thus obtain

αG − θ − arcsin

(DL√

(DLS)2 + (DS)2 tan2 βsin θ

)

+ arctan

(DS

DLStanβ

)= 0, (24)

where we used LR = DL. Eq. (24) is the gravitationallens equation, in the sense that it is an equation for thelensed image position θ when the intrinsic source positionβ and the angular distances DL, DS and DLS are given.We should stress that Eq. (24) is linear in αG. Thislinearity makes perturbative calculations much simpleras shown below.

Before going to iterative calculations, we mention a re-lation of Eq. (24) to an improved version of the gravita-tional lens equation by Bozza [20]. Eq. (24) is rearrangedas

αG − θ +B = arcsin

(DL√

(DLS)2 + (DS)2 tan2 βsin θ

),

(25)

where we used Eq. (22). By taking the sine of the bothsides of Eq. (25), we obtain

sinB cos(αG − θ) + cosB sin(αG − θ)

=DL√

(DLS)2 + (DS)2 tan2 βsin θ. (26)

By using Eqs. (19) and (23) for the triangle in Figure6, we obtain

sinB =DS tanβ√

(DLS)2 + (DS)2 tan2 β, (27)

cosB =DLS√

(DLS)2 + (DS)2 tan2 β. (28)

Substituting Eqs. (27) and (28) into Eq. (26) leads to

DS tanβ cos(αG − θ) +DLS sin(αG − θ) = DL sin θ.(29)

It is straightforward to rearrange Eq. (29) as

DS tanβ =DL sin θ −DLS sin(αG − θ))

cos(αG − θ). (30)

This is the improved expression of the lens equation byBozza [20]. We should note that this expression is highlynonlinear in αG. It seems that it is not suitable for iter-ative calculations in terms of a complicated form of αG,e.g. in modified gravity theories [17].

Before closing this section, we mention the strong de-flection, for which the light ray can have the windingnumber N ≥ 1. The photon trajectory is described bythe orbit equation (du/dφ)2 = F (u), where u = 1/r. Seee.g. Eq. (26) in Reference [11] for this equation. Theorbit equation gives the angle separation from the sourceto the receiver as

φRS = ±∫ R

S

du√F (u)

, (31)

where ± correspond to the anticlockwise or clockwise mo-tion, respectively. Therefore, φRS can be larger than 2π.As a result, also αI in Eq. (3) can. Corresponding tothis, αG has a modulo 2π, so that it can describe also thestrong deflection case. See Reference [12] for the strongdeflection in finite-distance cases. For the strong deflec-tion case, it should be noted that arcsin functions havea modulo 2π. Therefore, Eq. (24) is modified for thestrong deflection case as

αG − θ − arcsin

(DL√

(DLS)2 + (DS)2 tan2 βsin θ

)

+ arctan

(DS

DLStanβ

)= 2nπ, (32)

where n is an integer. Here, the sign of n is chosen asthe same as that of αG, such that n can mean the wind-ing number N of the light ray. Dabrowski and Schunckmentioned also the large deflection case, where they as-sume that the intersection point Q is in the lens plane[19]. Even if the intersection point Q is in the lens planeand the source and observer are at infinity, there are dif-ferences between their equation and the present resultsof Eqs. (24) and (32). This is because Dabrowski andSchunck equation is expressed in terms of distances be-tween objects instead of distances between planes.

6

III. ITERATIVE SOLUTIONS OF THEGRAVITATIONAL LENS EQUATION WITH

FINITE-DISTANCE EFFECTS

A. Iterative method for the finite-distancegravitational lens equation

Eq. (24) is the finite-distance gravitational lens equa-tion that holds for a general situation in a static andspherically symmetric spacetime. In this section, we shallexamine an iterative method for Eq. (24). For this pur-pose, we make an additional assumption that all the an-gles of β, θ and αG are small, namely |β| 1, |θ| 1and |αG| 1. Note that αG for a strong deflection casecan be larger than the order of unity, even if β and θ aresmall.

It is convenient to introduce a (nondimensional) book-keeping parameter ε in order to make the present iter-ative procedure more transparent. The intrinsic sourceposition β is given. Therefore, we do not expand β in ε.In the small angle approximation, β is small. Hence, itcan be expressed in terms of ε as

β = εβ(1). (33)

On the other hand, θ and αG are nonlinearly dependenton the intrinsic source position. In the small-angle ap-proximation, hence, they can be expressed in a Taylorseries as

θ =

∞∑k=1

εkθ(k), (34)

αG =

∞∑k=1

εkαG(k). (35)

Eqs. (33), (34) and (35) are substituted into Eq. (24).At the first order in ε, we obtain the linearized lens equa-tion

β(1) = θ(1) −DLS

DSαG(1). (36)

It seems that Eq. (36) is the same as the conventionallens equation. However, we should note that αG(1) in Eq.(36) contain effects of finite distances. For a given β, Eq.(36) is an equation for the unknown variable θ(1).

At the second order in ε, Eq. (24) becomes a linearequation for θ(2) and it is immediately solved as

θ(2) =DLS

DSαG(2). (37)

We thus obtain the second-order solution θ(2), becauseαG(2) is calculated by using θ(1).

At the third order in ε, Eq. (24) gives a solution for

θ(3).

θ(3) =DLS

DSαG(3)

+1

3

[1−

(DS

DLS

)2]

(β(1))3 +

1

2

DLDS

(DLS)2(β(1))

2θ(1)

+1

6

DL

DS

[1−

(DL

DLS

)2]

(θ(1))3. (38)

B. Einstein ring in an iterative scheme

The finite-distance effects on the deflection angle oflight are discussed in e.g. References [11, 17]. In theiriterative calculations for Schwarzschild, Kottler or Weylgravity models, the deflection angle at the lowest orderis 4m/b, where m is the lens mass. Eq. (9) is rewrittenas b = DL sin θ = εDLθ(1) + O(ε2). Eq. (35) in thesmall angle approximation means αG = O(ε). Therefore,the scaling of the lens mass should be m = ε2M , whereM ≡ m(2) and M is independent of ε. By substituting

b = DL sin θ and m = ε2M into the form of 4m/b, weobtain the linear order of αG in ε as

αG(1) =4M

DLθ(1), (39)

where we use θ(1) 6= 0.Only in this paragraph, we assume that the source is

located exactly behind the lens. Namely, β = 0 is as-sumed. By substituting Eq. (39) into Eq. (36), we obtaina quadratic equation as

(θ(1))2 =

4MDLS

DLDS. (40)

This means that the lensed image becomes a circle thatis usually called the Einstein ring. Therefore, we definethe radius of the Einstein ring by

θE(1) ≡√

4MDLS

DLDS. (41)

This definition is consistent with that in the conven-tional gravitational lens formulation that assumes thesmall angle approximation and the asymptotic receiverand source.

Rigorously speaking, the expression for the Einsteinring radius for exotic objects such as a wormhole [23, 24]may be different from Eq. (41) for the Schwarzschildspacetime or a spacetime model that approaches theSchwarzschild spacetime in a certain limit. See e.g. alsoEqs. (5) and (6) in Reference [25], in which Izumi et al.discussed the radius of the Einstein ring for an inversepower model (proposed by Kitamura et al. [26]) repre-senting the Ellis wormhole and Schwarzschild black holesin the weak field approximation.

7

We should note that the Einstein ring radius by Eq.(41) is valid only at the lowest order in iterative calcula-tions. The actual radius of the Einstein ring is dressed inthe present iteration scheme, because it is the sum of allthe terms in ε, namely

θE =

∞∑k=1

εkθE(k). (42)

The discussion and expressions in this section are gen-eral. In the Weyl conformal gravity case, θ(2) vanishes asshown in the next section.

IV. WEYL CONFORMAL GRAVITY ON THELENS EQUATION

A. Deflection of light in Weyl conformal gravity

The Weyl conformal gravity model was proposed byBach [27]. The action in the Weyl conformal gravity iswritten as

S =

∫d4x√−gCabcdCabcd, (43)

where g denotes the determinant of the metric. Birkoff’stheorem still holds even for a generalized solution in theWeyl conformal gravity [28]. The static and sphericallysymmetric vacuum solution in the Weyl conformal grav-ity was obtained by Mannheim and Kazanas [29]. Thissolution is expressed by using three new parameters (of-ten denoted as β, γ and k). It is written as

ds2 = −B(r)dt2 +B−1(r)dr2 + r2(dθ2 + sin2 θdφ2).(44)

B(r) can be approximated as

B(r) = 1− 3mγ − 2m

r+ γr − kr2. (45)

Here, mγ 1 is assumed, so that we can neglect m2γ(=m×mγ m). r2 terms have been already discussed inthe Kottler model [17]. For the simplicity, we ignore−kr2

in B(r) in the following. Mannheim and Kazanas arguedthat the Weyl gravity can explain the flat rotation ofgalaxies without introducing dark matter, for which γ isof the order of the inverse of the Hubble radius (denotedas rH), namely γ ∼ r−1

H [29]. We focus on this Weylgravity model. Physically, mγ 1 means that the blackhole under study is much smaller than the Hubble radiusof the present universe, namely m rH .

Takizawa et al. obtained the deflection angle of lightfor the receiver and source that are within a finite dis-tance from a lens object in Weyl conformal gravity [17].

It is

αWeyl =2m

b

(√1− b2u2

S +√

1− b2u2R

)−mγ

(buS√

1− b2u2S

+buR√

1− b2u2R

)+O

(m2, γ2

), (46)

where u defines the inverse distance as u ≡ 1/r, uR anduS denote the inverse distance from the lens to the re-ceiver and source, respectively. Note that terms linear inγ do not exist in αWeyl.

Several authors made attempts to calculate the deflec-tion angle in this spacetime in the literature [11, 30–34], though their discussions and methods are not self-consistent. For instance, they imagined the asymptoticreceiver and source in such an asymptotically nonflatspacetime. In another case, only the φRS was consid-ered.

By using Eq. (34) for Eq. (9), b is expanded in a seriesof ε as

b =DL sin θ,

=εDLθ(1) + ε2DLθ(2)

+ ε3DL

(θ(3) −

1

6(θ(1))

3

)+O(ε4). (47)

The parameter γ has no direct relation with ε. Con-sequently, γ = O(ε0). This is consistent with mγ 1,because mγ = O(ε2).

By using Eq. (47), we obtain

buR = εθ(1) +O(ε2), (48)

buS = εDL

DLSθ(1) +O(ε3), (49)

where DLS/LS = cos(π − φRS) = 1 + O(ε2) is used toobtain Eq. (49).

By substituting Eqs. (48) and (49) into Eq. (46), weobtain the Taylor series of αG in Eq. (35) up to the thirdorder as

αG(1) =DS

DLS

(θE(1))2

θ(1), (50)

αG(2) =− DS

DLS

(θE(1))2θ(2)

(θ(1))2, (51)

αG(3) =− DS

DLS

(θE(1))2

θ(1)

×[θ(3)

θ(1)−

(θ(2))2

(θ(1))2

+1

12

1 + 3

(DL

DLS

)2

+ 3γDS

DLSDL

(θ(1))

2

],

(52)

where we use Eqs. (34) and (41). Note the the parameterγ appears only through the last term of αG(3) in Eq. (52).

8

B. Lensed image positions in Weyl conformalgravity

By substituting Eq. (50) into Eq. (36), we obtain

β(1) = θ(1) −(θE(1))

2

θ(1), (53)

where θ(1) 6= 0. This equation is solved as

θ(1) =1

2

[β(1) ±

√(β(1))2 + 4(θE(1))2

]. (54)

This is in agreement with the know results in the con-ventional lens theory, which are corresponding to the so-called primary (or plus) and secondary (or minus) im-ages. However, Eq. (54) is still valid also for nonasymp-totic cases.

Next, we substitute Eq. (51) into Eq. (37). We imme-diately find the second-order solution as

θ(2) = 0. (55)

In the following, we thus use θ(2) = 0.Finally, we substitute Eq. (52) into Eq. (38) to obtain

a linear equation for θ(3). The solution for this equationis expressed in a long form as

θ(3) = θS(3) + θW(3), (56)

where θS(3) and θW(3) mean the third-order part only by the

lens mass (without γ) and that by γ parameter, respec-tively. They are

θS(3) =

(1 +

(θE(1))2

(θ(1))2

)−1

×

[− 1

12

1 + 3

(DL

DLS

)2

(θE(1))2θ(1)

+1

3

1−

(DS

DLS

)2

(β(1))3

+1

2

DLDS

(DLS)2(β(1))

2θ(1)

+1

6

DL

DS

1−

(DL

DLS

)2

(θ(1))3

], (57)

θW(3) =3γ

(1 +

(θE(1))2

(θ(1))2

)−1DLDS

DLS(θE(1))

2θ(1). (58)

V. OBSERVABILITY OF THE LENSED IMAGEPOSITION SHIFT DUE TO WEYL GRAVITY

The above calculations show that larger θE increasesthe third-order corrections including the Weyl gravity ef-fect. Therefore, we consider a cluster of galaxies as a lens

object in two cases separately. The first case is the so-called strong lensing, for which a lens system is close tothe Einstein ring. For instance, giant arcs are observednear a central part of a massive cluster of galaxies. Thesecond case is weak lensing, for which the source and thelens object are largely separated in the sky. This caseplays a role in cosmic shear measurements.

For the both cases, we consider a cluster of galaxieswith mass M ∼ 1014M. For its simplicity, we assumeDL ∼ DLS ∼ 1 Gpc, which means DS ∼ 2 Gpc. In thefollowing calculations, therefore, DL/DLS ∼ DS/DLS ∼DL/DS ∼ O(1). According to Eq. (41), the radius ofthe Einstein ring for this lens system becomes

θE(1) ∼ 10−4, (59)

which is corresponding to nearly one third arcminutes.The Weyl gravity model parameter γ is [29]

γ ∼ (rH)−1, (60)

where rH is the Hubble radius of the present universe,roughly speaking ∼ 10 Gpc.

A. Strong lensing case

We assume that a spherically symmetric lens systemis close to the Einstein ring, for which Eq. (46) can beused for describing the gravitational deflection of light.The system is nearly the Einstein ring, such that we canassume

β(1) θE(1), (61)

θ(1) ∼ θE(1). (62)

This means that

β(1) θE(1). (63)

By using these conditions for Eq. (57), we obtain

θS(3) ∼ (θE(1))3

∼ 10−12, (64)

where we used Eq. (59). This corresponds to O(10−1)microarcseconds. Hence, the third-order correction bythe finite-distance effects, which must exist also in thetheory of general relativity, is beyond reach of the currentVLBI technology.

Next, the third-order term purely in the Weyl confor-mal gravity is estimated as

θW(3) ∼ γDL(θE(1))3

∼ 10−13

(DL

1Gpc

)(10Gpc

rH

)(γ

(rH)−1

)(θE(1)

10−4

)3

,

(65)

9

where we used Eq. (59). This is ∼ 10 picoarcseconds, farbelow the current capability of EHT (∼ 30 microarcsec-onds).

By comparing Eqs. (64) and (65), we find the reasonwhy θW(3) is smaller by a factor of 10 than θS(3). θW(3) in-

cludes an extra factor DLγ ∼ DL/rH , which is ∼ 10−1 inthe above example. This implies that the more distant alens system is, the larger the Weyl gravity effect on thegravitational lens becomes.

B. Weak lensing case

As a second example, we consider weak lensing, forwhich

β(1) θE(1). (66)

As an example, we assume β(1) ∼ 10−3, which means theseparation angle ∼ 10 arcminutes for the above galaxycluster model. Then, the linear-order solution by Eq.(54) becomes

θp(1) ≡1

2

[β(1) +

√(β(1))2 + 4(θE(1))2

]∼ β(1), (67)

and

θm(1) ≡1

2

[β(1) −

√(β(1))2 + 4(θE(1))2

]∼ −β(1)

(θE(1)

β(1)

)2

. (68)

As a result, |θm(1)| θE(1) β(1) ∼ θp(1).

By using Eqs. (67) and (68) in Eq. (57), we obtain

θSp(3) ∼ (β(1))3

∼ (β(1))2θp(1), (69)

and

θSm(3) ∼ (β(1))3

(θE(1)

β(1)

)2

∼ θSp(3)

(θE(1)

β(1)

)2

. (70)

By comparing Eq. (69) with Eq. (70), we find θSm(3) is

much smaller by the factor (θE(1)/β(1))2 than θSp(3).

For the above galaxy cluster model in this section, Eqs.(69) and (70) are estimated as

θSp(3) ∼ 10−9

(β(1)

10−3

)3

, (71)

which is corresponding to ∼ 102 microarcseconds, and

θSm(3) ∼ 10−11

(θE(1)

10−4

)2( β(1)

10−3

), (72)

TABLE I. Summary of typical values for the strong and weaklensing by the Weyl gravity model, corresponding to SectionV. We assume a cluster of galaxies with ∼ 1014 solar masses at∼ 1 Gpc. Two cases are considered. One is the strong lensingas β(1) ∼ 0. The other is the weak lensing as β(1) ∼ 10−3.The angles in this table are in radians.

β(1) ∼ 0 ∼ 10−3

θE(1) 10−4 10−4

θp(1) 10−4 10−3

|θm(1)| 10−4 10−5

θSp(3) 10−12 10−9

|θSm(3) | 10−12 10−11

θWp(3) 10−13 10−12

|θWm(3) | 10−13 10−14

which is corresponding to ∼ one microarcsecond. Thelatter value is beyond the current capability, while theformer one is corresponding to ∼ 0.1 milliarcsecondswhich are larger than the current VLBI accuracy. There-

fore, the third-order effect as θSp(3) can be relevant with

VLBI observations. On the other hand, it is difficult to

detect effects by θSp(3) through weak lensing observations

by optical telescopes that have currently the best imagequality of ∼ 0.1 arcseconds (∼ 102 milliarcseconds).

Next, we examine the third-order correction purely bythe Weyl conformal gravity. Eq. (58) for the primaryimage θp(1) becomes

θWp(3) ∼ γDL(θE(1))

2β(1)

∼ 10−12

(DL

1Gpc

)(10Gpc

rH

)(γ

(rH)−1

)×(θE(1)

10−4

)2( β(1)

10−3

), (73)

which is corresponding to ∼ 10−1 microarcseconds. Thisis far below the current EHT accuracy (∼ 30 microarc-seconds). As a result, effects of the Weyl gravity modelare negligible in the current and near-future lensing ob-servations.

VI. CONCLUSION

We discussed the finite-distance lens equation thatis consistent with the deflection angle Takizawa et al.define[17]. The present lens equation, though it is equiv-alent to the lens equation by Bozza [20], is linear in thedeflection angle and therefore it makes iterative calcula-tions much simpler,

As an explicit example of an asymptotically nonflatspacetime, we considered a static and spherically sym-metric solution in Weyl conformal gravity. We focused

10

on the Weyl gravity model relevant with the flat rotationof galaxies, for which γ parameter in the Weyl gravitymodel is of the order of the inverse of the present Hubbleradius [29]. For this case, we obtained iterative solutionsfor the finite-distance lens equation up to the third or-der. The effect of the Weyl gravity on the lensed imageposition begins at the third order and it is linear in theimpact parameter of light.

The deviation of the lensed image position from thegeneral relativistic one is ∼ 10−2 microarcsecond for thelens and source with a separation angle of ∼ 1 arcminute,where we consider a cluster of galaxies with 1014M at∼ 1 Gpc for instance. The deviation becomes ∼ 10−1

microarcseconds, even if the separation angle is ∼ 10 ar-cminutes. Therefore, effects of the Weyl gravity modelare negligible in current and near-future observations ofgravitational lensing. On the other hand, the general rel-ativistic corrections at the third order ∼ 0.1 milliarcsec-onds can be relevant with VLBI observations. However,the discussions in this paper are limited within a spher-

ically symmetric model. Asymmetric cases are an openissue. Further study along this direction is left for future.

ACKNOWLEDGMENTS

We are grateful to Marcus Werner for the useful discus-sions. We wish to thank Emanuel Gallo for the helpfulcomments on his recent works with his collaborators. Wewould like to thank Mareki Honma for the conversationson the EHT method and technology. We thank YuuitiSendouda, Ryuichi Takahashi, Masumi Kasai, Kei Ya-mada, Ryunosuke Kotaki, Masashi Shinoda, and HideakiSuzuki for the useful conversations. This work was sup-ported in part by Japan Society for the Promotion ofScience (JSPS) Grant-in-Aid for Scientific Research, No.18J14865 (T.O.), No. 20K03963 (H.A.), in part by Min-istry of Education, Culture, Sports, Science, and Tech-nology, No. 17H06359 (H.A.) and in part by JSPS re-search fellowship for young researchers (T.O.).

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