1. Introduction

The asymptotic structure of spacetime and the Newman-Penrose (NP) formalism¹ are powerful tools for analyzing the behavior of isolated systems and compact sources in General Relativity (GR). This formalism has found many applications in the last past years in different areas of the GR, and also in numerical relativity. From the mathematical point of view, the NP approach is a special case of the tetrad formalism, where a set of four null vectors is chosen to form a basis at each point of the spacetime. Then, several complex scalars are introduced to describe the dynamic of the gravitational fields at infinity, and the evolution equations of these complex scalars, i.e. the Bianchi identities, are written. Furthermore, these vectors are appropriate for studying the properties of congruences of null geodesics², such congruences naturally arise in the context of the propagation of the gravitational and the electromagnetic radiation³.

As has been mentioned before, the NP formalism has been used in several aspects of analytical and computational relativity: recently, in collaboration with Kozameh, this formalism was used to describe the behavior of compact sources, such as Black Holes collisions and close binary coalescences, linking the dynamics of the system to the gravitational radiation emitted. Moreover, we find some similarities and several differences with other approaches like the Post-Newtonian equations⁴. Furthermore, the NP approach is used for extracting gravitational waves from numerical simulation via the computation of one of the five Weyl complex scalars. Although the NP formalism is primarily focused on studying the properties of radiation of isolated systems, this formulation is a very useful framework for constructing and investigating exact solutions of the four-dimensional General Relativity. Additionally, there are attempts to generalize this approach to higher dimensions⁵. Particularly, the method is very powerful when the spacetime is algebraically special according the Petrov classification⁶, since some complex scalars can be set to zero by choosing an appropriate tetrad aligned with the outgoing tangent vector fields of null radial geodesics.

In this article, we propose to make a short review about the NP approach, giving some applications of this approach and explaining the physical interpretation of the most important scalars introduced by Newman and Penrose. This article is organized as follows: In Sec. 2 we give the geometric notion of an asymptotically flat spacetime, and we introduce a set of null vectors to define the complex Weyl scalars, the Maxwell scalars, and the twelve spin coefficients used in the NP formalism. In the Sec. 3, we reduce all the equations and definitions to a particular set of coordinates, usually called Bondi coordinates, and we introduce the notion of the Mass and the Bondi linear momentum. The BMS group and the transformation between different families of null cuts is shown in Sec. 4, while the physical meaning of the Weyl scalars is discussed in Secs. 5. In Sec. 6 , we use the NP approach to analyze a more general stationary axisymmetric spacetime, where we compute all the NP quantities, and we apply the resulting equations to some familiar spacetimes. Finally, we conclude this work giving some finals remarks and comments.

2. Asymptotically flat spacetime

The notion of asymptotically flat spacetime is an adequate tool to analyze the gravitational and electromagnetic radiation coming from an arbitrary compact source. During the 60s, Bondi, Sachs and collaborators^7-9 used a system of canonical coordinates to define the mass, momentum and gravitational radiation. Then, Penrose defines the notion of asymptotically flat spaces (or asymptotically simple), using the idea of re-scale the metric by an appropriate factor, usually called conformal factor, which is appropriately chosen to decay to zero at infinity¹⁰. The geometric notion of Penrose can be summarized in the following definition

Definition: A spacetime (\scrm, 𝑔 𝑎𝑏 ) is called asymptotically simple, if the curvature tensor goes to zero as we approach to infinity in the future direction of the null geodesics of the spacetime. These geodesics end up in what is called the future null infinity. A future null asymptote is a manifold M with boundary M together with a smooth Lorentzian metric 𝑔 𝑎𝑏 , and a smooth function 𝛺 on $\hat\scrm$ satisfying the following

M = M

On M, 𝑔 𝑎𝑏 = 𝛺 2 𝑔 𝑎𝑏 with 𝛺>0

At 𝛺=0, 𝑛 𝑎 ≡ 𝜕 𝑎 𝛺≠0 and 𝑔 𝑎𝑏 𝑛 𝑎 𝑛 𝑏 =0

From this geometric definition, it is possible to show that,

If M satisfy the vacuum Einstein equations near $\scri$ then $\scri$ it is a null boundary. Also $\scri$ consists of two disjoint parts $\scri^+$ and $\scri^-$, each topologically 𝑆 2 ×𝑅.

The Weyl tensor 𝐶 𝑎𝑏𝑐 𝑑 vanishes at $\scri$ and the peeling assumption [8] establishes the way it approaches to zero. Furthermore, the Weyl tensor is conformally invariant, i.e. the tensor 𝐶 𝑎𝑏𝑐 𝑑 constructed using the re-scaled metric 𝑔 𝑎𝑏 and 𝐶 𝑎𝑏𝑐 𝑑 are the same at null infinity¹¹.

We begin the study of the null infinity properties, introducing a coordinate system in the neighborhood of M, which we will label as (u,r,ζ, ζ )¹². In this system, the time u represents a family of null surfaces, r is the affine parameter along the null geodesics of the constant u surfaces, and ζ= e iϕ cot(θ/2), the complex stereographic angle labeling the null geodesics of M. We reach taking r→∞, thus the null infinity is described by the remaining coordinates (u,ζ, ζ ). Now, the two-surface metric becomes,

making the usual choice of 𝛺= 𝑟 −1 as the conformal factor, Eq. ([metricasphy]) gives the induced metric on,

Here 𝑃(𝑢,𝜁, 𝜁 ) is a strictly positive function which depends on the framework choice.

3. Bondi coordinates

As mentioned in Sec. 2.1, it is possible to introduce a general set of coordinates in the neighborhood of $\scri^+$ since the conformal factor 𝑃, in Eq. ([metrica]), provides a great freedom when they are chosen. However, in many practical applications, it is useful to restrict the transformation imposing the condition

With this choice, the two-surface metric ([metrica]) becomes a sphere. These coordinates are then called Bondi coordinates. In this section, we will give the main equations of the NP formalism written in the Bondi system. These particular systems will correspond to inertial frames in General Relativity. However, the choice of the Bondi coordinate system is not unique, the coordinate transformations between two Bondi systems is called the Bondi-Metzner-Sachs (BMS) transformations¹⁴. Now, making the choice of the 𝑃 factor by imposing Eq. (35), and since 𝑃 =0, one can reduce the equations introduced in Sec. 2.2 as follows

Also the “eth” operators ð, ð can be written as

In many applications, it is quite convenient to define the so called Mass Aspect Ψ from the Weyl scalar ψ 2 0 ^7,

which satisfies the reality condition Ψ= Ψ . Finally the Bianchi identities in the Bondi system take the form,

Also, it is possible to write Eq. ([psi2dot]) in terms of 𝛹 as follows,

In the same way, the SC can be written in terms of the shear σ 0 and the Weyl scalars ψ n 0 .

4. The BMS group

The set of coordinate transformations at $\scri^+$ preserving the conditions in the Bondi coordinates is called the Bondi-Metzner-Sachs Group (BMS)^8,14,15. This group is the same as the asymptotic symmetry group that arises from the infinitesimal generators, i.e. from the asymptotic Killing vectors. Now, to construct the BMS group, we start considering the following mapping

where 𝑢,𝜁, and 𝑟 are the standard Bondi-type coordinates. Under this coordinate transformation all the relations developed up to this point are preserved. The fraction linear transformations ([anguloBMS]) are the only one-to-one mapping of the sphere to itself. Then, the BMS group is defined by the mapping ([anguloBMS]) and the following restriction on the transformations ([nuBMS0]),

where 𝛼(𝜁, 𝜁 ) is a regular arbitrary function on the sphere. Now, it is possible to show from Eq. ([nuBMS]) and ([anguloBMS]) that the spherical metric transforms as¹⁶,

where the conformal factor 𝐾 associated with this transformation is given by

and where

Now, the infinite-parameter subgroup obtained by setting 𝐾=1,

is known as the supertranslation subgroup. A supertranslation α(ζ, ζ ) moves points on each generator by an amount α(ζ, ζ ). This function can be expressed in terms of infinite constants using, for example, a tensorial spin-s harmonic expansion¹⁷ in the following way,

In this expansion, α 0 , and α i represents the ordinary translations. For extra details about the Lorentz group and the BMS transformations the reader could see Refs.^16,18.

5. Physical interpretation of the asymptotic scalars

In this section, we will discuss the physical interpretation of some of the complex scalars used in the NP formalism. A simpler way to introduce this topic, is to focus on the dominant terms of the peeling expansion, and make a tensorial spin-s harmonic expansion of these functions. The most important scalars in our analysis are the following,

For any asymptotically flat axially symmetric spacetime, the Komar integral¹⁹ gives a precise notion of the angular momentum. The Komar formula uses the Killing vector field to define the global angular momentum. Assuming that the axis of symmetry is labeled as z-axis, the non-zero component of the angular momentum can be written as follows [20],

Since the real contribution of [ σ 0 ð σ 0 ] i is zero for any axisymmetric spacetimes. One only need to use the real part of ψ 1 0i to define the dipole mass moment as follows,

Now, if the spacetime has no global symmetries, the mass dipole-angular momentum two form will add some extra contributions of the free data σ 0 and its derivatives. Notably, as one can see in the literature, there are many definitions of the angular momentum for isolated systems in general relativity. A recent living review²¹ offers a complete survey of the main results in the field with the main motivations and technical aspects of each definition, the fact that there is no agreement among these alternative approaches reflects the difficulty of the subject. Although, in a recent work together with Kozameh⁴, we introduced a new definition of dipole mass moment-angular momentum tensor using the Winicour-Tamburino linkage²²,

This definition allows to define the center of mass and write the equation of motion of the center of mass linking the time evolution with the emitted gravitational radiation. On the other hand, as we discuss in Sec. 3, for a Bondi system, there is a precise notion of mass and linear momentum. Now, we can use the ℓ=0,1 components of the tensorial expansion previously introduced, and relate 𝜓 2 0 to the Bondi 4-momentum (𝑀, 𝑃 𝑖 ), as shown in the following equations,

The ℓ≥2 terms of the l.h.s of Eqs. (79) or (80), are the so called “supermomentum” at null infinity. Now, in the NP approach, the scalar 𝛹 4 measures the gravitational radiation received at null infinity. This scalar is related to the gravitational wave modes as follows,

where ℎ + , ℎ × are the plus mode, and cross mode of the gravitational wave in the transverse traceless gauge²³, respectively. Thus, the complex function 𝜎 0 , introduced in Eq. ([radiation]), yields the gravitational radiation reaching at null infinity, and 𝜎 𝑅 𝑖𝑗 = ℎ + 𝑖𝑗 , and 𝜎 𝐼 𝑖𝑗 = ℎ × 𝑖𝑗 are the quadrupolar contributions of the gravitational wave.

Finally, we focus on the Maxwell scalars, which give information about the electromagnetic contribution received at $\scri^+$. At this point, we can mention that the electric charge is the zeroth order in the tensorial expansion of ϕ 1 0 , and the dipole electromagnetic moment corresponds to the ℓ=1 component of ϕ 0 0 ^3,24, i.e,

where 𝑝 𝑖 , 𝜇 𝑖 are the electric and magnetic dipole moment respectively.

6. Application to stationary axisymmetric spacetimes

Stationary axisymmetric spacetimes are of great importance in General Relativity, astrophysics, Newtonian gravity, and also in Post-Newtonian theories. Many sources like stars, galaxies, accretion disks, and black holes are modeled under these assumptions of temporal and axial symmetry. These global symmetries play a central role in analytic calculations, since they are very useful when the field equations are simplified. In this section, we will focus on studying these kinds of spacetimes. Now, following Ref. 25, we can introduce a

more general stationary axisymmetric metric in the standard spherical coordinates (𝑡,𝑟,𝜃,𝜑) as follows,

where the metric functions 𝜇 0 , 𝜇 1 , 𝜇 2 , 𝜇 3 , and 𝜔 are arbitrary functions of (𝑟,𝜃). The stationary and axisymmetric character of (84), is reflected in the fact that the metric coefficients are independent of the coordinate 𝑡, and the azimuthal angle 𝜑, and also that the spacetime is invariant under simultaneous transformations 𝑡→−𝑡 and 𝜑→−𝜑. Now, we are considering spacetimes which are asymptotically flat in the neighborhood of infinity, thus to ensure an adequate behaviour of the line element ([gabS]), the metric functions 𝜇 0 , 𝜇 1 , 𝜇 2 , 𝜇 3 , and the angular velocity 𝜔 must go to zero as 𝑟→∞. Also we choose the signature (+,−,−,−) in agreement with the orthonormal conditions introduced in Sec. 2.1.

One of the main ingredients in the NP formalism is the construction of a complex null tetrad ( 𝑙 𝑎 , 𝑛 𝑎 , 𝑚 𝑎 , 𝑚 𝑎 ). For that, we introduce first an orthogonal tetrad constructed from the timelike foliation 𝑡=const, then the normal vector to the hypersurface 𝛴 𝑡 is given by

Now, on 𝛴 𝑡 we can find three spacelike vectors denoted by

Also, the set of vectors ( 𝑡 𝑎 , 𝑟 𝑎 , 𝑒 𝜃 𝑎 , 𝑒 𝜑 𝑎 ) satisfy

Now, from these vectors we can build a null tetrad making the following linear combinations,

then, our null tetrad is given by the null vectors,

and lowering the indices using 𝑔 𝑎𝑏 we find the conjugate tetrad which is given by

Finally, we will assume the following potential vector

where 𝜒, 𝐴 𝑟 , 𝐴 𝜃 , and 𝐴 𝜑 are also functions of (𝑟,𝜃). Now, using the null tetrad, and the potential vector 𝐴 𝑎 , we can compute all the quantities introduced in the previous sections to solve this axially symmetric stationary metric in the NP formalism. In the appendix A, we will show the general set of equations, but in the next three subsections, we will reduce the metric given by Eq. (84) to the Reissner-Nordström, Chazy-Curzon, and Kerr spacetimes, by solving an algebraic system of five equations for the metric functions of (84), and we will write all the complex scalars on the basis given by Eqs. (94-97).