# Representation theory of the Lorentz group (for undergraduate students of physics)

The Lorentz group is a Lie group of symmetries of the spacetime of special relativity. This group can be realized as a collection of matrices, linear transformations, or unitary operators on some Hilbert space; it has a variety of representations. In any relativistically invariant physical theory, these representations must enter in some fashion; physics itself must be made out of them. Indeed, special relativity together with quantum mechanics are the two physical theories that are most thoroughly established, and the conjunction of these two theories is the study of the infinite-dimensional unitary representations of the Lorentz group. These have both historical importance in mainstream physics, as well as connections to more speculative present-day theories.
The full theory of the finite-dimensional representations of the Lie algebra of the Lorentz group is deduced using the general framework of the representation theory of semisimple Lie algebras. The finite-dimensional representations of the connected component SO(3; 1)$^+$ of the full Lorentz group O(3; 1) are obtained by employing the Lie correspondence and the matrix exponential. The full finite-dimensional representation theory of the universal covering group (and also the spin group, a double cover) SL(2, C) of SO(3; 1)$^+$ is obtained, and explicitly given in terms of action on a function space in representations of SL(2, C) and sl(2, C). The representatives of time reversal and space inversion are given in space inversion and time reversal, completing the finite-dimensional theory for the full Lorentz group. The general properties of the (m, n) representations are outlined. Action on function spaces is considered, with the action on spherical harmonics and the Riemann P-functions appearing as examples. The infinite-dimensional case of irreducible unitary representations is classified and realized for Lie algebras. Finally, the Plancherel formula for SL(2, C) is given.
The development of the representation theory has historically followed the development of the more general theory of representation theory of semisimple groups, largely due to Élie Cartan and Hermann Weyl, but the Lorentz group has also received special attention due to its importance in physics. Notable contributors are physicist E. P. Wigner and mathematician Valentine Bargmann with their Bargmann–Wigner programme, one conclusion of which is, roughly, a classification of all unitary representations of the inhomogeneous Lorentz group amounts to a classification of all possible relativistic wave equations. The classification of the irreducible infinite-dimensional representations of the Lorentz group was established by Paul Dirac´s doctoral student in theoretical physics, Harish-Chandra, later turned mathematician, in 1947.
The non-technical introduction contains some prerequisite material for readers not familiar with representation theory. The Lie algebra basis and other adopted conventions are given in conventions and Lie algebra bases.

## Non-technical introduction to representation theory

The present purpose is to illustrate the role of representation theory of groups in mathematics and in physics. Rigor and detail take the back seat, as the main objective is to fix the notion of finite-dimensional and infinite-dimensional representations of the Lorentz group. The reader familiar with these concepts should skip by.

### Symmetry groups

The mathematical notion of a group and the notion of symmetry in both mathematics and physics are intimately related. A group has the simple property that if one element of a group is multiplied by another, the result is another element of the group. The same can, mutatis mutandis, be said of symmetries. Apply one symmetry operation (physically or by changing coordinate system), and then another one. The result is that of applying a single symmetry operation. Else they don't qualify as a symmetry operations the present context. Group theory is thus the mathematical language in which symmetries of nature are expressed. These may relate to very concrete symmetries of physical objects, like the symmetries of a square. One then speaks of the symmetry group associated with the object.
In the case of a square, the symmetry group, called the dihedral group D$_4$, is finite. For instance, only some rotations, and some reflections in the plane, will make the transformed square look exactly like it did before the symmetry operation. Other objects possess higher symmetry. The sphere is the extreme example. It possesses full rotational symmetry and reflectional symmetry. Rotate or reflect a ball with any kind of rotation or reflection about any plane through the origin, and it will look exactly the same as before the symmetry operation.
A central fact is that the symmetry groups can be represented by matrices. In the case of D$_4$ for the square, the matrix representation is composed of eight 2 × 2 matrices. In the case of the symmetries of a sphere, the matrix group is the orthogonal group of three dimensions. These are 3 × 3 matrices.

### Symmetry of space and time

Less obvious is that space itself possesses symmetry. It too looks the same, no matter how one rotates it, so it has rotational symmetry. In fact, in this case, it is more practical to use passive rotations, meaning the observer rotates himself and does not attempt to physically rotate the universe. Mathematically, the active operation of a rotation is performed by multiplying position vectors by a rotation matrix. A passive rotation is accomplished by rotating only the basis vectors of the coordinate system. (Envisage the coordinate system being fixed in the rotated observer. Then actively rotate the observer only.) In this way, every point in space obtains new coordinates, just as if it was somehow physically rigidly rotated. The Lorentz group contains all rotation matrices, extended to four dimensions with zeros in the first row and the first column except for the upper left element which is one, as elements. There are, in addition, matrices that effect Lorentz boosts. These can be thought of in the passive view as (instantly!) giving the coordinate system (and with it the observer) a velocity in a chosen direction. Two special transformations are used to invert the coordinate system in space, space inversion, and in time, time reversal. In the first case, the space coordinate axes are reversed. The latter is reversal of the time direction. This is best though of as just having the observer set his clock at minus what it shows and then have the clock's hands move counterclockwise. Physical time progresses forward as always.

### Lorentz transformations

In the spacetime of special relativity, called Minkowski space, space and time are interwoven. Thus the four coordinates of points in spacetime, called events, change in ways unexpected before the advent of special relativity, with time dilation and length contraction as two immediate consequences. The four-dimensional matrices of Lorentz transformations compose the Lorentz group. Its elements represent symmetries, and just like physical objects can be rotated using rotation matrices, the same physical objects (whose coordinates now include the time coordinate) can be transformed using the matrices representing Lorentz transformations. In particular, the four-vector representing an event in Lorentz frame transforms as
    $x'={\begin{bmatrix}x'^{0}\\x'^{1}\\x'^{2}\\x'^{3}\end{bmatrix}}={\begin{bmatrix}\lambda _{00}&\lambda _{01}&\lambda _{02}&\lambda _{03}\\\lambda _{10}&\lambda _{11}&\lambda _{12}&\lambda _{13}\\\lambda _{20}&\lambda _{21}&\lambda _{22}&\lambda _{23}\\\lambda _{30}&\lambda _{31}&\lambda _{32}&\lambda _{33}\end{bmatrix}}{\begin{bmatrix}x^{0}\\x^{1}\\x^{2}\\x^{3}\end{bmatrix}},$
or on short form
    $x'=\Lambda x.$

### Multiplication table and representations

The basic feature of every finite group is its multiplication table, also called Cayley table, that records the result of multiplying any two elements. A representation of a group can be thought of new set of elements, finite-dimensional or infinite-dimensional matrices, giving the same multiplication table after mapping the old elements to the new elements in a one-to-one fashion. The same holds true in the case of an infinite group like the rotation group SO(3) or the Lorentz group. The multiplication table is just harder to visualize in the case of a group of uncountable size (same size as the set of reals).

### Ordinary Lorentz transformations matrices do not suffice

The objects to be transformed may be something else than ordinary physical objects extending in three spatial dimensions (and time, unless the frame is the rest frame). It is at this point that representation theory enters the picture. The electromagnetic field is usually envisaged by assignment to each point in space a three-dimensional vector representing the electric field and another three-dimensional vector representing the magnetic field. When space is rotated, the expected thing happens. The electric field and the magnetic field vectors at a designated point rotate with preserved length and angle between them. Under Lorentz boosts they behave differently, and in a way showing that the two vectors certainly aren't separate physical objects. The electric and magnetic components mix. See the illustration on the right. The electromagnetic field tensor displays the manifestly covariant mathematical structure of the electromagnetic field.

### Finite-dimensional representations by matrices

The problem of representation theory of the Lorentz group is, in the finite-dimensional case, to find new sets of matrices, not necessarily 4 × 4 in size that satisfies the same multiplication table as the matrices in the original Lorentz group. Returning to the example of the electromagnetic field, what is needed here are 6 × 6-matrices that can be applied to a 6-dimensional column vector containing the all together six components of the electromagnetic field. Thus one is looking for 6 × 6-matrices such that
    $F'={\begin{bmatrix}E'^{1}\\E'^{2}\\E'^{3}\\B'^{1}\\B'^{1}\\B'^{3}\end{bmatrix}}={\begin{bmatrix}\Pi (\Lambda )_{00}&\Pi (\Lambda )_{01}&\Pi (\Lambda )_{02}&\Pi (\Lambda )_{03}&\Pi (\Lambda )_{04}&\Pi (\Lambda )_{05}\\\Pi (\Lambda )_{10}&\Pi (\Lambda )_{11}&\Pi (\Lambda )_{12}&\Pi (\Lambda )_{13}&\Pi (\Lambda )_{14}&\Pi (\Lambda )_{15}\\\Pi (\Lambda )_{20}&\Pi (\Lambda )_{21}&\Pi (\Lambda )_{22}&\Pi (\Lambda )_{23}&\Pi (\Lambda )_{24}&\Pi (\Lambda )_{25}\\\Pi (\Lambda )_{30}&\Pi (\Lambda )_{31}&\Pi (\Lambda )_{32}&\Pi (\Lambda )_{33}&\Pi (\Lambda )_{34}&\Pi (\Lambda )_{35}\\\Pi (\Lambda )_{40}&\Pi (\Lambda )_{41}&\Pi (\Lambda )_{42}&\Pi (\Lambda )_{43}&\Pi (\Lambda )_{44}&\Pi (\Lambda )_{45}\\\Pi (\Lambda )_{50}&\Pi (\Lambda )_{51}&\Pi (\Lambda )_{52}&\Pi (\Lambda )_{53}&\Pi (\Lambda )_{54}&\Pi (\Lambda )_{55}\end{bmatrix}}{\begin{bmatrix}E^{1}\\E^{2}\\E^{3}\\B^{1}\\B^{1}\\B^{3}\end{bmatrix}},$
in short
    $F'=\Pi (\Lambda )F,$
correctly expresses the transformation of the electromagnetic field under the Lorentz transformation Λ. The same reasoning can be applied to Dirac's bispinors. While these have 4 components, the original 4 × 4-matrices in the Lorentz group will not do the job properly, not even when restricted to mere rotations. Another 4 × 4-representation is needed.
The sections dedicated to finite-dimensional representations are dedicated to exposing all such representations by finite-dimensional matrices that respect the multiplication table.

### Infinite-dimensional representations by action on vector spaces of functions

Infinite-dimensional representations are usually realized as acting on sets of real or complex-valued functions on a set X endowed with a group action. A set being endowed with a group action A means, in essence, that if x ∈ X and g ∈ G that A(g)x = y with y ∈ X. Now if C$^X$ denotes the set of all complex-valued functions on X, which is a vector space, a representation Π of G can be defined by
    $(\Pi (g))f(x)=f(A(g^{-1})x),\quad f\in \mathbf {C} ^{X},g\in G,x\in X.$
The point to make is that again one has
    $(\Pi (g))f\in \mathbf {C} ^{X}.$
and one has a representation of G. This representation of G is finite-dimensional if and only if X is a finite set. This method is very general, and one typically explores vector spaces of more specialized functions on sets close at hand. To illustrate this procedure, consider a group G of n-dimensional matrices as a subset of Euclidean space R$^{n2}$, and let the space of functions be polynomials, perhaps of some maximum degree d, or even homogeneous polynomials of degree d, all defined on R$^{n2}$. Then restrict those functions to G ⊂ R$^{n2}$. Now observe that the set X = G automatically comes equipped with group actions, namely
    $L_{g}h=gh,R_{g}h=hg^{-1},C_{g}h=ghg^{-1},g,h\in G.$
Here L$_g$ denotes left action (by g), R$_g$ denotes right action (by g), and C$_g$ denotes conjugation (by g). With this sort of action, the vectors being acted on are functions. The resulting representations are (when the functions are unrestricted), in the first and second cases respectively, the left regular representation and the right regular representation of G on C$^G$.
The goal in the infinite-dimensional case of the representation theory is to classify all different possible representations, and to exhibit them in terms of vector spaces of functions and the action of the standard representation on the arguments of the functions.

### Infinite-dimensional representations viewed as infinite-dimensional matrices

In order to relate this to the finite-dimensional case, one may chose a basis for the vector space of functions and simply then examine what happens to the basis functions under a given transformation. Take image of the first basis function under a transformation, expressed as a linear combination the basis functions. Explicitly, if f$_1$, f$_2$, ... is a basis, compute
    ${\begin{aligned}\Pi (\Lambda )f_{1}&=\lambda _{11}f_{1}+\lambda _{21}f_{2}+\cdots ,\\\Pi (\Lambda )f_{2}&=\lambda _{12}f_{1}+\lambda _{22}f_{2}+\cdots ,\\&\,\,\,\vdots \end{aligned}}$
The coefficients of the basis functions in this expression is then the first column in a representative matrix. Proceed. In general, the resulting matrix is countably infinite in dimension:
    $\Pi (\Lambda )={\begin{bmatrix}\lambda _{11}&\lambda _{12}&\cdots \\\lambda _{21}&\lambda _{22}&\cdots \\\vdots &\vdots &\ddots \end{bmatrix}}$
Again, it is required that the set of infinite matrices obtained this way stand in one-to-one correspondence with the original 4 × 4-matrices and that the multiplication table is the right one - the one of the 4 × 4-matrices. It should be emphasized that in the infinite-dimensional case, one is rarely concerned with these matrices. They are exposed here only to highlight the common thread. But individual matrix elements are frequently computed, especially for the Lie algebra (below).

### Lie algebra

The Lorentz group is a Lie group and has as such a Lie algebra, The Lie algebra is a vector space of matrices that can be said to model the group near the identity. It is endowed with a multiplication operation, the Lie bracket. With it, the product in the group can near the identity be expressed in Lie algebraic terms (but not in a particularly simple way). The link between the (matrix) Lie algebra and the (matrix) Lie group is the matrix exponential. It is one-to-one near the identity in the group.
Due to this it often suffices to find representations of the Lie algebra. Lie algebras are much simpler objects than Lie groups to work with. Due to the fact that the Lie algebra is a finite-dimensional vector space, in the case of the Lorentz Lie algebra the dimension is 6, one need only find a finite number of representative matrices of the Lie algebra, one for each element of a basis of the Lie algebra as a vector space. The rest follow from extension by linearity, and the representation of the group is obtained by exponentiation.
The metric signature to be used below is (−1,  1 ,  1,  1) and the metric is given by η = diag(−1,  1,  1,  1). The physics convention for Lie algebras and the exponential mapping is used. These choices are arbitrary, but once they are made, fixed. One possible choice of basis for the Lie algebra is, in the standard representation, given by
    ${\begin{aligned}J_{1}&=J^{23}=-J^{32}=i{\begin{pmatrix}0&0&0&0\\0&0&0&0\\0&0&0&-1\\0&0&1&0\\\end{pmatrix}},\\[5pt]J_{2}&=J^{31}=-J^{13}=i{\begin{pmatrix}0&0&0&0\\0&0&0&1\\0&0&0&0\\0&-1&0&0\\\end{pmatrix}},\\[5pt]J_{3}&=J^{12}=-J^{21}=i{\begin{pmatrix}0&0&0&0\\0&0&-1&0\\0&1&0&0\\0&0&0&0\\\end{pmatrix}},\\[5pt]K_{1}&=J^{01}=J^{10}=i{\begin{pmatrix}0&1&0&0\\1&0&0&0\\0&0&0&0\\0&0&0&0\\\end{pmatrix}},\\[5pt]K_{2}&=J^{02}=J^{20}=i{\begin{pmatrix}0&0&1&0\\0&0&0&0\\1&0&0&0\\0&0&0&0\\\end{pmatrix}},\\[5pt]K_{3}&=J^{03}=J^{30}=i{\begin{pmatrix}0&0&0&1\\0&0&0&0\\0&0&0&0\\1&0&0&0\\\end{pmatrix}}.\end{aligned}}$

## Applications

Many of the representations, both finite-dimensional and infinite-dimensional, are important in theoretical physics. Representations appear in the description of fields in classical field theory, most importantly the electromagnetic field, and of particles in relativistic quantum mechanics, as well as of both particles and quantum fields in quantum field theory and of various objects in string theory and beyond. The representation theory also provides the theoretical ground for the concept of spin. The theory enters into general relativity in the sense that in small enough regions of spacetime, physics is that of special relativity.
The finite-dimensional irreducible non-unitary representations together with the irreducible infinite-dimensional unitary representations of the inhomogeneous Lorentz group, the Poincare group, are the representations that have direct physical relevance.
Infinite-dimensional unitary representations of the Lorentz group appear by restriction of the irreducible infinite-dimensional unitary representations of the Poincaré group acting on the Hilbert spaces of relativistic quantum mechanics and quantum field theory. But these are also of mathematical interest and of potential direct physical relevance in other roles than that of a mere restriction. There were speculative theories, (tensors and spinors have infinite counterparts in the expansors of Dirac and the expinors of Harish-Chandra) consistent with relativity and quantum mechanics, but they have found no proven physical application. Modern speculative theories potentially have similar ingredients per below.

### Mathematics

From the point of view that the goal of mathematics is to classify and characterize, the representation theory of the Lorentz group is since 1947 a finished chapter. But in association with the Bargmann–Wigner programme, there are (as of 2006) yet unresolved purely mathematical problems, linked to the infinite-dimensional unitary representations.
The irreducible infinite-dimensional unitary representations may have indirect relevance to physical reality in speculative modern theories since the (generalized) Lorentz group appears as the little group of the Poincare group of spacelike vectors in higher spacetime dimension. The corresponding infinite-dimensional unitary representations of the (generalized) Poincaré group are the so-called tachyonic representations. Tachyons appear in the spectrum of bosonic strings and are associated with instability of the vacuum. Even though tachyons may not be realized in nature, these representations must be mathematically understood in order to understand string theory. This is so since tachyon states turn out to appear in superstring theories too in attempts to create realistic models.
One open problem (as of 2006) is the completion of the Bargmann–Wigner programme for the isometry group SO(D – 2, 1) of the de Sitter spacetime dS$_{D–2}$. Ideally, the physical components of wave functions would be realized on the hyperboloid dS$_{D–2}$ of radius μ > 0 embedded in R$^{D−2,1}$ and the corresponding O(D − 2, 1) covariant wave equations of the infinite-dimensional unitary representation to be known.
It is common in mathematics to regard the Lorentz group to be, foremost, the Möbius group to which it is isomorphic. The group may be represented in terms of a set of functions defined on the Riemann sphere. These are the Riemann P-functions, which are expressible as hypergeometric functions.

### Classical field theory

While the electromagnetic field together with the gravitational field are the only classical fields providing accurate descriptions of nature, other types of classical fields are important too. In the approach to quantum field theory (QFT) referred to as second quantization, the starting point is one or more classical fields, where e.g. the wave functions solving the Dirac equation are considered as classical fields prior to (second) quantization. While second quantization and the Lagrangian formalism associated with it is not a fundamental aspect of QFT, it is the case that so far all quantum field theories can be approached this way, including the standard model. In these cases, there are classical versions of the field equations following from the Euler–Lagrange equations derived from the Lagrangian using the principle of least action. These field equations must be relativistically invariant, and their solutions (which will qualify as relativistic wave functions according to the definition below) must transform under some representation of the Lorentz group.
The action of the Lorentz group on the space of field configurations (a field configuration is the spacetime history of a particular solution, e.g. the electromagnetic field in all of space over all time is one field configuration) resembles the action on the Hilbert spaces of quantum mechanics, except that the commutator brackets are replaced by field theoretical Poisson brackets.

### Relativistic quantum mechanics

For the present purposes the following definition is made: A relativistic wave function is a set of n functions ψ$^α$ on spacetime which transforms under an arbitrary proper Lorentz transformation Λ as
    $\psi '^{\alpha }(x)=D{[\Lambda ]^{\alpha }}_{\beta }\psi ^{\beta }(\Lambda ^{-1}x),$
where D[Λ] is an n-dimensional matrix representative of Λ belonging to some direct sum of the (m, n) representations to be introduced below.
The most useful relativistic quantum mechanics one-particle theories (there are no fully consistent such theories) are the Klein–Gordon equation and the Dirac equation in their original setting. They are relativistically invariant and their solutions transform under the Lorentz group as Lorentz scalars ((m, n) = (0, 0)) and bispinors respectively ((0, ​$^1$⁄$_2$) ⊕ (​$^1$⁄$_2$), 0)). The electromagnetic field is a relativistic wave function according to this definition, transforming under (1, 0) ⊕ (0, 1).

### Quantum field theory

In QFT, the demand for relativistic invariance enters, among other ways in that the S-matrix necessarily must be Poincaré invariant. This has the implication that there is one or more infinite-dimensional representation of the Lorentz group acting on Fock space. One way to guarantee the existence of such representations is the existence of a Lagrangian description (with modest requirements imposed, see the reference) of the system using the canonical formalism, from which a realization of the generators of the Lorentz group may be deduced.
The transformations of field operators illustrate the complementary role played by the finite-dimensional representations of the Lorentz group and the infinite-dimensional unitary representations of the Poincare group, witnessing the deep unity between mathematics and physics. For illustration, consider the definition of some n-component field operator: Given a matrix representation as above, a relativistic field operator is a set of n operator valued functions on spacetime which transforms under proper Lorentz transformations Λ according to
    $\Psi ^{\alpha }(x)\rightarrow \Psi '^{\alpha }(x')=U[\Lambda ]\Psi ^{\alpha }(x)U[\Lambda ^{-1}]=D{[\Lambda ^{-1}]^{\alpha }}_{\beta }\Psi ^{\beta }(\Lambda x)$
Here U[Λ] is the unitary operator representing Λ on the Hilbert space on which Ψ is defined. By considerations of differential constraints that the field operator must be subjected to in order to describe a single particle with definite mass m and spin s (or helicity), it is deduced that
    $\Psi ^{\alpha }(x)=\sum _{\sigma }\int dp[a(\mathbf {p} ,\sigma )u^{\alpha }(\mathbf {p} ,\sigma )e^{ip\cdot x}+a^{\dagger }(\mathbf {p} ,\sigma )v^{\alpha }(\mathbf {p} ,\sigma )e^{-ip\cdot x}],$         (X1)
where a$^†$, a are interpreted as creation and annihilation operators respectively. The creation operator a$^†$ transforms according to
    $a^{\dagger }(\mathbf {p} ,\sigma )\rightarrow a'^{\dagger }(\mathbf {p} ',\sigma )=U[\Lambda ]a^{\dagger }(\mathbf {p} ,\sigma )U[\Lambda ^{-1}]=a^{\dagger }(\Lambda \mathbf {p} ,\rho )D^{(s)}{[R(\Lambda ,\mathbf {p} )^{-1}]^{\rho }}_{\sigma },$
and similarly for the annihilation operator. The point to be made is that the field operator transforms according to a finite-dimensional non-unitary representation of the Lorentz group, while the creation operator transforms under the infinite-dimensional unitary representation of the Poincare group characterized by the mass and spin (m, s) of the particle. The connection between the two are the wave functions, also called cofficient functions
    $u^{\alpha }(\mathbf {p} ,\sigma )e^{ip\cdot x},\quad v^{\alpha }(\mathbf {p} ,\sigma )e^{-ip\cdot x}$
that carry both the indices (x, α) operated on by Lorentz transformations and the indices (p, σ) operated on by Poincaré transformations. This may be called the Lorentz–Poincaré connection. Too exhibit the connection, subject both sides of equation (X1) to a Lorentz transformation resulting in for e.g. u,
    ${D[\Lambda ]^{\alpha }}_{\alpha '}u^{\alpha '}(\mathbf {p} ,\lambda )={D^{(s)}[R(\Lambda ,\mathbf {p} )]^{\lambda '}}_{\lambda }u^{\alpha }(\Lambda \mathbf {p} ,\lambda '),$
where D is the non-unitary Lorentz group representative of Λ and D$^{(s)}$ is a unitary representative of the so-called Wigner rotation R associated to Λ and p that derives from the representation of the Poincaré group, and s is the spin of the particle.
All of the above formulas, including the definition of the field operator in terms of creation and annihilation operators, as well as the differential equations satisfied by the field operator for a particle with specified mass, spin and the (m, n) representation under which it is supposed to transform, and also that of the wave function, can be derived from group theoretical considerations alone once the frameworks of quantum mechanics and special relativity is given.

### Speculative theories

In theories in which spacetime can have more than D = 4 dimensions, the generalized Lorentz groups O(D − 1; 1) of the appropriate dimension take the place of O(3; 1).
The requirement of Lorentz invariance takes on perhaps its most dramatic effect in string theory. Classical relativistic strings can be handled in the Lagrangian framework by using the Nambu–Goto action. This results in a relativistically invariant theory in any spacetime dimension. But as it turns out, the theory of open and closed bosonic strings (the simplest string theory) is impossible to quantize in such a way that the Lorentz group is represented on the space of states (a Hilbert space) unless the dimension of spacetime is 26. The corresponding result for superstring theory is again deduced demanding Lorentz invariance, but now with supersymmetry. In these theories the Poincaré algebra is replaced by a supersymmetry algebra which is a Z$_2$-graded Lie algebra extending the Poincaré algebra. The structure of such an algebra is to a large degree fixed by the demands of Lorentz invariance. In particular, the fermionic operators (grade 1) belong to a (0, 1/2) or (1/2, 0) representation space of the (ordinary) Lorentz Lie algebra. The only possible dimension of spacetime in such theories is 10.

## Finite-dimensional representations

Representation theory of groups in general, and Lie groups in particular, is a very rich subject. The full Lorentz group is no exception. The Lorentz group has some properties that makes it "agreeable" and others that make it "not very agreeable" within the context of representation theory. The group is simple and thus semisimple, but is not connected, and none of its components are simply connected. Perhaps most importantly, the Lorentz group is not compact.
For finite-dimensional representations, the presence of semisimplicity means that the Lorentz group can be dealt with the same way as other semisimple groups using a well-developed theory. In addition, all representations are built from the irreducible ones, since the Lie algebra possesses the complete reducibility property. But, the non-compactness of the Lorentz group, in combination with lack of simple connectedness, cannot be dealt with in all the aspects as in the simple framework that applies to simply connected, compact groups. Non-compactness implies, for a connected simple Lie group, that no nontrivial finite-dimensional unitary representations exist. Lack of simple connectedness gives rise to spin representations of the group. The non-connectedness means that, for representations of the full Lorentz group, time reversal and space inversion has to dealt with separately.

### History

The development of the finite-dimensional representation theory of the Lorentz group mostly follows that of the subject in general. Lie theory originated with Sophus Lie in 1873. By 1888 the classification of simple Lie algebras was essentially completed by Wilhelm Killing. In 1913 the theorem of highest weight for representations of simple Lie algebras, the path that will be followed here, was completed by Élie Cartan. Richard Brauer was 1935–38 largely responsible for the development of the Weyl-Brauer matrices describing how spin representations of the Lorentz Lie algebra can be embedded in Clifford algebras. The Lorentz group has also historically received special attention in representation theory, see History of infinite-dimensional unitary representations below, due to its exceptional importance in physics. Mathematicians Hermann Weyl and Harish-Chandra and physicists Eugene Wigner and Valentine Bargmann made substantial contributions both to general representation theory and in particular to the Lorentz group. Physicist Paul Dirac was perhaps the first to manifestly knit everything together in a practical application of major lasting importance with the Dirac equation in 1928.

### Strategy

Classification of the finite-dimensional irreducible representations of a semisimple Lie algebra g over R or C generally consists of two steps. The first step amounts to analysis of hypothesized representations resulting in a tentative classification. The second step is actual realization of these representations.
A real Lie algebra is usually complexified enabling analysis in an algebraically closed field. Working over the complex numbers in addition admits nicer bases. The following theorem applies: A real linear finite-dimensional representations of a real Lie algebra extends to a complex linear representations of its complexification. The real linear representation is irreducible if and only if the corresponding complex linear representation is irreducible. Moreover, a complex semisimple Lie algebra has the complete reducibility property. This means that every finite-dimensional representation decomposes as a direct sum of irreducible representations.
Classification thus amounts to studying irreducible complex linear representations of the (complexified) Lie algebra.

#### Step one

The first step is to hypothesize the existence of irreducible representations. The properties of these hypothetical representations are investigated, primarily by using the Lie algebra.
Conditions on a sets of properties of an irreducible representations necessary for the existence of an irreducible representation are established. The sets of properties in question are called weights. In the simplest description, a weight λ for a representation (π, V) is a simultaneous eigenvalue (λ(H$_1$), λ(H$_2$), ..., λ(H$_n$)) for the representatives under π of the elements of an ordered basis (H$_1$, H$_2$, ..., H$_n$) of a Cartan subalgebra h ⊂ g. A Cartan subalgebra is, for complex semisimple Lie algebras, a maximal commutative subalgebra such that each ad$_H$, where ad$_H$ is the representative of H in the adjoint representation is diagonalizable.
A partial ordering on the set of weights is defined, and the notion of highest weight in terms of this partial ordering is established for any set of wheights. Using the structure on the Lie algebra, the notions dominant weight and integral weight are defined. A representation (π, V) is called highest weight cyclic if it contains a non-zero vector v such that π(H$_i$)v = λ(H$_i$)v for all i, and if V is generated by π(g)v. In addition, no strictly higher weights than λ may occur. Highest weight cyclic representations are irreducible and highest weight cyclic representations with the same highest weight are isomormhic.

##### Result

An irreducible representation is highest weight cyclic with a dominant integral highest weight. This is part of Cartan's theorem of highest weight. The full theorem of the highest weigh contains step two below as well. In summary, it provides a classification of the irreducible representations in terms of the weights of the Lie algebra.

#### Step two

Step two is composed of showing the following: Each dominant integral weight of a complex semisimple Lie algebra gives rise to a highest weight cyclic representation.
Step one has the side benefit that the structure of the irreducible representations is better understood. Representations decompose as direct sums of weight spaces, with the weight space corresponding to the highest weight one-dimensional. Repeated application of the representatives of certain elements of the Lie algebra called lowering operators yields a set of generators for the representation as a vector space. The application of one such operator on a vector with definite weight results either in zero or a vector with strictly lower weight. Raising operators work similarly, but results in a vector with strictly higher weight or zero. The representatives of the Cartan subalgebra acts diagonally in a basis of weight vectors.
There are several standard ways of constructing irreducible representations:
* Construction using Verma modules. This approach is purely Lie algebraic. (Generally applicable to complex semisimple Lie algebras.)
* Construction using the Peter–Weyl theorem, also called the compact group approach. (Generally applicable to complex semisimple Lie algebras.)
* Construction using the Borel–Weil theorem, in which holomorphic representations of the group G corresponding to g are contructed. (Generally applicable to complex semisimple Lie algebras.)
* Performing standard operations on known representations, in particular applying Clebsch–Gordan decomosition to tensor products of representations. (Not generally applicable.)
* In the simplest cases, construction from scratch

##### Result

The composite result of the two steps is:
    Each dominant integral weight of a complex semisimple Lie algebra gives rise to an irreducuble highest weight cyclic representation. These are the only irreducible representations.
All items in the above list work in the case of sl(2, C). What will actually be used below are irreducible representations of SL(2, C) constructed from scratch, yielding via (C1) irreducible representations of sl(2, C) with highest weight 0, 1, ... and their complex conjugate representations.

#### Unitarian trick

For some semisimple Lie algebras, especially non-compact ones, it is easier to proceed indirectly via Weyl's unitarian trick instead of applying Cartan's theorem directly. In the present case of so(3; 1) a chain of isomorphisms between Lie algebras and other correspondences preserving irreducible representations is set up, so that the representations may be obtained from representations of SU(2) ⊗ SU(2). See equation (A1) and references around it. It is essential here that SU(2) is compact, since then the irreducible representations of SU(2) ⊗ SU(2) are simply tensor products of irreducible representations of SU(2), that can all be obtained from the irreducible representations of su(2).
Cartan's theorem is applied to sl(2, C), together with knowledge of its highest weights, yields a classification of the representations of so(3; 1) via (A1). An explicit construction of the irreducible representations of SL(2, C) is presented, yielding explicit irreducible representations of sl(2, C), thus, via (A1), completing the task with the (m, n) representations of so(3; 1) as the final result.
Representative matrices may be obtained by choice of basis in the representation space. An explicit formula for matrix elements is presented and some common representations are listed.

#### Lie group representations

The Lie correspondence is subsequently employed for obtaining group representations of the connected component of the Lorentz group, SO(3, 1)$^+$. This is effected by taking the matrix exponential of the matrices of the Lie algebra representation. A subtlety arises due to the (in physics parlance) doubly connected nature of SO(3, 1)$^+$. This results in the projective representations or two-value representations that are actually spin representations of the covering group SL(2, C).
The Lie correspondence gives results only for the connected component of the groups, and thus the components of the full Lorentz that contain the operations of time reversal and space inversion are treated separately, mostly from physical considerations, by defining representatives for the space inversion and time reversal matrices.

### The Lie algebra

According to the strategy, the irreducible complex linear representations of the complexification, so(3; 1)$_C$ of the Lie algebra so(3; 1) of the Lorentz group are to be found. A convenient basis for so(3; 1) is given by the three generators J$_i$ of rotations and the three generators K$_i$ of boosts. They are explicitly given in conventions and Lie algebra bases.
The Lie algebra is complexified, and the basis is changed to the components of
    $\mathbf {A} ={\frac {\mathbf {J} +i\mathbf {K} }{2}}\,,\quad \mathbf {B} ={\frac {\mathbf {J} -i\mathbf {K} }{2}}\,.$
The components of A = (A$_1$, A$_2$, A$_3$) and B = (B$_1$, B$_2$, B$_3$) separately satisfy the commutation relations of the Lie algebra su(2) and, moreover, they commute with each other,
    $\left[A_{i},A_{j}\right]=i\varepsilon _{ijk}A_{k}\,,\quad \left[B_{i},B_{j}\right]=i\varepsilon _{ijk}B_{k}\,,\quad \left[A_{i},B_{j}\right]=0,$
where i, j, k are indices which each take values 1, 2, 3, and ε$_{ijk}$ is the three-dimensional Levi-Civita symbol. Let A$_C$ and B$_C$ denote the complex linear span of A and B respectively.
One has the isomorphisms Template:Equation box 1
where sl(2, C) is the complexification of su(2) ≈ A ≈ B.
The utility of these isomorphisms comes from the fact that all irreducible representations of su(2), and hence (see strategy) all irreducible complex linear representations of sl(2, C), are known. According to the final conclusion in strategy, the irreducible complex linear representation of sl(2, C) is isomorphic to one of the highest weight representations. These are explicitly given in complex linear representations of sl(2, C).

#### The unitarian trick

In (A1), all isomorphisms are C-linear (the last is just a defining equality). The most important part of the manipulations below is that the R-linear (irreducible) representations of a (real or complex) Lie algebra are in one-to-one correspondence with C-linear (irreducible) representation of its complexification (see Strategy). With this in mind, it is seen that the R-linear representations of the real forms of the far left, so(3; 1), and the far right, sl(2, C), in (A1) can be obtained from the C-linear representations of sl(2, C) ⊕ sl(2, C).
The manipulations to obtain representations of a non-compact algebra (here so(3; 1)), and subsequently the non-compact group itself, from qualitative knowledge about unitary representations of a compact group (here SU(2)) is a variant of Weyl's so-called unitarian trick. The trick specialized to SL(2, C) can be summarized concisely.
Let V be a finite-dimensional complex vector space. The objects in the following list are in one-to-one correspondence. The correspondence is given via complexification of Lie algebras, via restriction to real forms, via the exponential mapping (to be introduced), and via a standard mechanism (also to be introduced) for obtaining Lie algebra representations given group representations:
* Representations of SU(2) on V.
* Holomorphic representations of SL(2, C) on V.
* Representations of su(2) on V.
* Complex linear representations of sl(2, C) on V.
In this list, direct products (groups) or direct sums (Lie algebras) may be introduced if done consistently across per below. The essence of the trick is that the starting point in the above list is immaterial. Both qualitative knowledge (like existence theorems for one item on the list or properties like irreducibility) and concrete realizations for one item on the list will translate and propagate, respectively, to the others.
The representations of sl(2, C) ⊕ sl(2, C), which is the Lie algebra of SL(2, C) × SL(2, C), are according to the strategy required to be irreducible. This means that they must be tensor products of complex linear representations of sl(2, C), as can be seen by restriction to the subgroup SU(2) × SU(2) ⊂ SL(2, C) × SL(2, C), a compact group to which the Peter–Weyl theorem applies. and hence orthonormality of irreducible characters may be appealed to. The irreducible unitary representations of SU(2) × SU(2) are precisely the tensor products of irreducible unitary representations of SU(2). These stand in one-to-one correspondence with the holomorphic representations of SL(2, C) × SL(2, C) and these, in turn, are in one-to-one correspondence with the complex linear representations of sl(2, C) ⊕ sl(2, C) because SL(2, C) × SL(2, C) is simply connected. A new list, directly applicable to (A1) is obtained. The following objects are in one-to-one correspondence:
* Representations of SU(2) × SU(2) on V.
* Holomorphic representations of SL(2, C) × SL(2, C) on V.
* Representations of su(2) ⊕ su(2) on V.
* Complex linear representations of sl(2, C) ⊕ sl(2, C) on V.
The highest weight representations of sl(2, C) are indexed by μ for μ = 0, 1/2, 1,  … . (The highest weights are actually 2μ = 0, 1, 2, ..., but see below for the notation.) The tensor products of two complex linear factors then form the irreducible complex linear representations of sl(2, C) ⊕ sl(2, C). For reference, if (π$_1$, U) and (π$_2$, V) are representations of a Lie algebra g, then their tensor product (π$_1$ ⊗ π$_2$, U ⊗ V) is given by either of
    ${\begin{aligned}\pi _{1}\otimes \pi _{2}(X)&=\pi _{1}(X)\otimes \mathrm {Id} _{V}+\mathrm {Id} _{U}\otimes \pi _{2}(X),\quad X\in \mathbf {g} ,\\\qquad \pi _{1}\otimes \pi _{2}(X,Y)&=\pi _{1}(X)\otimes \mathrm {Id} _{V}+\mathrm {Id} _{U}\otimes \pi _{2}(Y),\quad (X,Y)\in \mathbf {g} \oplus \mathbf {g} ,\end{aligned}}$         (A0)
where Id is the identity operator. Here, the latter interpretation is intended. All irreducible real linear, but not necessarily complex linear, representations of sl(2, C) follow from the last inclusion in (A1).

#### The representations

The representations for all Lie algebras and groups involved in the unitarian trick can now be obtained. The real linear representations for sl(2, C) and so(3; 1) follow here assuming the complex linear representations of sl(2, C) are known. Explicit realizations and group representations are given later.

##### sl(2, C)

The complex linear representations of the complexification of sl(2, C), sl(2, C)$_C$, obtained via isomorphisms in (A1), stand in one-to-one correspondence with the real linear representations of sl(2, C). The set of all, at least real linear, irreducible representations of sl(2, C) are thus indexed by a pair (μ, ν). The complex linear ones, corresponding precisely to the complexification of the real linear su(2) representations, are of the form (μ, 0), while the conjugate linear ones are the (0, ν). All others are real linear only. The linearity properties follow from the canonical injection, the far right in (A1), of sl(2, C) into its complexification. Representations on the form (ν, ν) or (μ, ν) ⊕ (ν, μ) are given by real matrices (the latter is not irreducible). Explicitly, the real linear (μ, ν)-representations of sl(2, C) are
Template:Equation box 1
where Φ$_μ$, μ = 0, 1/2, 1, 3/2, … are the complex linear irreducible representations of sl(2, C) and Φ$_ν$, ν = 0, 1/2, 1, 3/2, … their complex conjugate representations. (The labeling is sometimes in the literature 0, 1, 2, … or 0, 2, 4, …, but half-integers are chosen here to conform with the labeling for the so(3, 1) Lie algebra.) Here the tensor product is interpreted in the former sense of (A0). These representations are concretely realized below.

##### so(3; 1)

Via the displayed isomorphisms in (A1) and knowledge of the complex linear irreducible representations of sl(2, C) ⊕ sl(2, C), upon solving for J and K, all irreducible representations of so(3; 1)$_C$, and, by restriction, those of so(3; 1) are known. It's worth noting that the representations of so(3; 1) obtained this way are real linear (and not complex or conjugate linear) because the algebra is not closed upon conjugation, but they are still irreducible. Since so(3; 1) is semisimple, all its representations can be built up as direct sums of the irreducible ones.
Thus the finite dimensional irreducible representations of the Lorentz algebra are classified by an ordered pair of half-integers m = μ and n = ν, conventionally written as one of
    $(m,n)\equiv D^{(m,n)}\equiv \pi _{m,n}.$
The notation D$^{(m,n)}$ is usually reserved for the group representations. Let π$_{(m,n)}$ : so(3; 1) → gl(V), where V is a vector space, denote the irreducible representations of so(3; 1) according to this classification. These are, up to a similarity transformation, uniquely given by
Template:Equation box 1
where the J$^{(n)}$ = (J$^{(n)}$$_1$, J$^{(n)}$$_2$, J$^{(n)}$$_3$) are the (2n + 1)-dimensional irreducible spin n representations of so(3) ≈ su(2) and 1$_n$ is the n-dimensional unit matrix.

###### Explicit formula for matrix elements

Let π$_{(m,n)}$ : so(3; 1) → gl(V), where V is a vector space, denote the irreducible representations of so(3; 1) according to the (m,  n) classification. In components, with −m ≤ a, a′ ≤ m, −n ≤ b, b′ ≤ n, the representations are given by
Template:Equation box 1
where δ is the Kronecker delta and the J$_i$$^{(n)}$ are the (2n + 1)-dimensional irreducible representations of so(3), also termed spin matrices or angular momentum matrices. These are explicitly given as
Template:Equation box 1

###### Common representations

| | m=0 | 1/2 | 1 |
| n=0 | scalar | Weyl spinorbispinor | self-dual 2-form2-form field |
| 1/2 | Weyl spinor (right-handed) | 4-vector | Rarita–Schwinger field |
| 1 | anti-self-dual 2-form | | traceless symmetric tensor |
| Purple: (m, n) complex irreducible representations  Black: (m, n) ⊕ (n, m) Bold: (m, m) ||||
Since for any irreducible representation for which m ≠ n it is essential to operate over the field of complex numbers, the direct sum of representations (m, n) and (n, m) has a particular relevance to physics, since it permits to use linear operators over real numbers.
* (0, 0) is the Lorentz scalar representation. This representation is carried by relativistic scalar field theories.
* (1/2, 0) is the left-handed Weyl spinor and (0, 1/2) is the right-handed Weyl spinor representation. Fermionic supersymmetry generators transform under one of these representations.
* (1/2, 0) ⊕ (0, 1/2) is the bispinor representation. (See also Dirac spinor and Weyl spinors and bispinors below.)
* (1/2, 1/2) is the four-vector representation. The four-momentum of a particle (either massless or massive) transforms under this representation.
* (1, 0) is the self-dual 2-form field representation and (0, 1) is the anti-self-dual 2-form field representation.
* (1, 0) ⊕ (0, 1) is the representation of a parity-invariant 2-form field (a.k.a. curvature form). The electromagnetic field tensor transforms under this representation.
* (1, 1/2) ⊕ (1/2, 1) is the Rarita–Schwinger field representation.
* (1, 1) is the spin 2 representation of a traceless symmetric tensor field. A physical example is the traceless part of the energy-momentum tensor T$^{μν}$.
* (3/2, 0) ⊕ (0, 3/2) would be the symmetry of the hypothesized gravitino. It can be obtained from the (1, 1/2) ⊕ (1/2, 1)-representation.

### The group

The approach in this section is based on theorems that, in turn, are based on the fundamental Lie correspondence. The Lie correspondence is in essence a dictionary between connected Lie groups and Lie algebras. The link between them is the exponential mapping from the Lie algebra to the Lie group, denoted exp:g → G. It is one-to-one in a neighborhood of the identity.

#### The Lie correspondence

The Lie correspondence and some results based on it needed here and below are stated for reference. If G denotes a linear Lie group (i.e. a group of matrices) and g a linear Lie algebra (again a set of matrices), let Γ(g) denote the group generated by exp(g), the image of the Lie algebra under the exponential mapping (which is the Matrix exponential in this case), and let L(G) denote the Lie algebra of G (interpreted as the set of matrices X such that e$^{itX}$ ∈ G for all t ∈ R). The Lie correspondence reads in modern language, here specialized to linear Lie groups, as follows:
* There is a one-to-one correspondence between connected linear Lie groups and linear Lie algebras given by G ↔ g with g = L(G) or, equivalently G = Γ(g), expressed as Γ(L(G)) = G, respectively L(Γ(g)) = g. Lie
The following are some corollaries that will be used in the sequel:
   * A connected linear Lie group G is abelian if and only if g is abelian. Lie i
   * A connected subgroup H with Lie algebra h of a connected linear Lie group G is normal if and only if h ⊂ g is an ideal. Lie ii
   * If G, H are linear Lie groups with Lie algebras g, h and Π:G → H is a group homomorphism, then π:g → h, its pushforward at the identity, is a Lie algebra homomorphism and Π(e$^{iX}$) = e$^{iπ(X)}$ for every X ∈ g. Lie iii

#### Lie algebra representations from group representations

Using the above theorem it is always possible to pass from a representation of a Lie group G to a representation of its Lie algebra g. If Π : G → GL(V) is a group representation for some vector space V, then its pushforward (differential) at the identity, or Lie map, π : g → End V is a Lie algebra representation. It is explicitly computed using
    $\pi (X)={\frac {d}{dt}}\Pi (e^{itX})|_{t=0},\quad X\in {\mathfrak {g}}.$         (G6)
This, of course, holds for the Lorentz group in particular, but not all Lie algebra representations arise this way because their corresponding group representations may not exist as proper representations, i.e. they are projective, see below.

#### Group representations from Lie algebra representations

Given a so(3; 1) representation, one may try to construct a representation of SO(3; 1)$^+$, the identity component of the Lorentz group, by using the exponential mapping. Since SO(3; 1)$^+$ is a matrix Lie group, the exponential mapping is simply the matrix exponential. If X is an element of so(3; 1) in the standard representation, then
    $\Lambda =e^{iX}\equiv \sum _{n=0}^{\infty }{\frac {(iX)^{n}}{n!}}$         (G1)
is a Lorentz transformation by general properties of Lie algebras. Motivated by this and the Lie correspondence theorem stated above, let π : so(3; 1) → gl(V) for some vector space V be a representation and tentatively define a representation Π of SO(3; 1)$^+$ by first setting
    $\Pi _{U}(e^{iX})=e^{i\pi (X)},\quad X\in {\mathfrak {so}}(3;1).$         (G2)
The subscript U indicates a small open set containing the identity. Its precise meaning is defined below. There are at least two potential problems with this definition. The first is that it is not obvious that this yields a group homomorphism, or even a well defined map at all (local existence). The second problem is that for a given g ∈ U ⊂ SO(3; 1)$^+$ there may not be exactly one X ∈ so(3; 1) such that g = e$^{iX}$ (local uniqueness). The soundness of the tentative definition (G2) is shown in several steps below:
* Π$_U$ is a local homomorphism.
* Π(g) defined along a path using properties of Π$_U$ is a global homomorphism.
* The exponential mapping exp:so(3; 1) → SO(3; 1)$^+$ is surjective.
* Π(g) defined along a path coincides with Π$_U$(g) with U = SO(3; 1)$^+$.

##### Local existence and uniqueness

A theorem based on the inverse function theorem states that the map exp : so(3; 1) → SO(3; 1)$^+$ is one-to-one for X small enough A. This makes the map well-defined. The qualitative form of the Baker–Campbell–Hausdorff formula then guarantees that it is a group homomorphism, still for X small enough B. Let U ⊂ SO(3; 1)$^+$ denote image under the exponential mapping of an open set in so(3; 1) where conditions (A) and (B) both hold. Let g, h ∈ U, g = e$^X$, h = e$^Y$, then
    ${\begin{aligned}\Pi _{U}(gh)&=\Pi _{U}(e^{X}e^{Y})=e^{i\pi \log(e^{X}e^{Y})}&{\text{(by definition and by }}{\textbf {(A)}}{\text}}\\&=e^{i\pi (X+Y+{\frac {1}{2}}[X,Y]+{\frac {1}{12}}[X,[X,Y]]-{\frac {1}{12}}[Y,[X,Y]]+\ldots )}&{\text{(by Baker–Campbell–Hausdorff and }}{\textbf {(B)}}{\text}}\\&=e^{i\pi (X)+i\pi (Y)+{\frac {1}{2}}[i\pi (X),i\pi (Y)]+{\frac {1}{12}}[i\pi (X),[i\pi (X),i\pi (Y)]]-{\frac {1}{12}}[i\pi (Y),[i\pi (X),i\pi (Y)]]+\ldots )}&{\text{(since }}\pi {\text{ is a Lie algebra homomorphism)}}\\&=e^{\log(e^{i\pi (X)}e^{i\pi (Y)})}&{\text{(by Baker–Campbell–Hausdorff and }}{\textbf {(B)}}{\text{ again)}}\\&=e^{\log(\Pi _{U}(g)\Pi _{U}(h))}=\Pi _{U}(g)\Pi _{U}(h)&{\text{(by definition and by }}{\textbf {(A)}}{\text}}\\\end{aligned}}$         (G7)
This shows that the map Π$_U$ is a well-defined group homomorphism on U.

##### Global existence and uniqueness

Technically, formula (G2) is used to define Π near the identity. For other elements g ∉ U a path is chosen from the identity to g and defines Π along that path by partitioning it finely enough so that formula (G2) can be used again on the resulting factors in the partition. In detail,
    $g=g_{n}=(g_{n}g_{n-1}^{-1})(g_{n-1}g_{n-2}^{-1})\cdots (g_{2}g_{1}^{-1})(g_{1}g_{0}),\qquad \Pi (g)\equiv \Pi _{U}(g_{n}g_{n-1}^{-1})\Pi _{U}(g_{n-1}g_{n-2}^{-1})\cdots \Pi _{U}(g_{2}g_{1}^{-1})\Pi _{U}(g_{1}g_{0}),\quad g_{0}=1$         (G3)
where the g$_i$ are on the path and the factors on the far right are uniquely defined by (G2) provided that all g$_i$ g$_{i+1}$$^{−1}$ ∈ U and, for all conceivable pairs h,k of points on the path between g$_i$ and g$_{i+1}$, hk$^{−1}$ ∈ U as well. For each i take, by the inverse function theorem, the unique X$_i$ such that exp(X$_i$) = g$_i$g$_{i−1}$$^{−1}$ and obtain
    $\Pi (g)=\Pi _{U}(e^{iX_{n}})\Pi _{U}(e^{iX_{n-1}})\cdots \Pi _{U}(e^{iX_{2}})\Pi _{U}(e^{iX_{1}})=e^{i\pi (X_{n})}e^{i\pi (X_{n-1})}\cdots e^{i\pi (X_{2})}e^{i\pi (X_{1})}.$         (G4)
By compactness of the path there is an n large enough so that Π(g) is well defined, possibly depending on the partition and/or the path, whether g is close to the identity or not.

###### Partition independence

It turns out that the result is always independent of the partitioning of the path. To demonstrate the independence of a chosen path, the Baker–Campbell–Hausdorff formula is employed. It shows that Π$_U$ is a group homomorphism for elements in U.
To see this, first fix a partitioning used in (G3). Then insert a new point h somewhere on the path, say
    $g=\cdots (g_{i+1}h^{-1})(hg_{i}^{-1})\cdots ,\qquad \cdots \Pi _{U}(g_{i+1}h^{-1})\Pi _{U}(hg_{i}^{-1})\cdots .$
But
    $\cdots \Pi _{U}(g_{i+1}h^{-1})\Pi _{U}(hg_{i}^{-1})\cdots =\cdots \Pi _{U}(g_{i+1}h^{-1}hg_{i}^{-1})\cdots =\cdots \Pi _{U}(g_{i+1}g_{i}^{-1})\cdots$
as a consequence of the Baker–Campbell–Hausdorff formula and the conditions on the original partitioning. Thus, adding a point on the path has no effect on the definition of Π(g).
Then, for any two given partitions of a given path, they have common refinement, their union. This refinement can be reached from any of the two partitionings by, one-by-one, adding points from the other partition. No individual addition changes the definition of Π(g), hence, since there are finitely many points in each partition, the value of Π(g) must have been the same for the two partitionings to begin with.

###### Path independence

For simply connected groups, the construction will be independent of the path as well, yielding a well defined representation. In that case formula (G2) can unambiguously be used directly. Simply connected spaces have the property that any two paths can be continuously deformed into each other. Any such deformation is called a homotopy and is usually chosen as a continuous function H from the unit square {s,t ∈ R: 0 ≤ s, t ≤ 1} into the group. For s = 0 the image is one of the paths, for s = 1 the other, for intermediate s, an intermediate path results, but endpoints are kept fixed.
One deforms the path, a little bit at a time, using the previous result, the independence of partitioning. Each consecutive deformation is so small that two consecutive deformed paths can be partitioned using the same partition points. Thus two consecutive deformed paths yield the same value for Π(g). But any two pairs of consecutive deformations need not have the same choice partition points, so the actual path laid out in the group as progression is made through the deformation does indeed change.
Using compactness arguments, in a finite number of steps, the original (s = 0) path is deformed into the other (s = 1) without affecting the value of Π(g).

###### Global homomorphism

The map Π$_U$ is, by the Baker-Campbell-Hausdorff formula, a local homomorphism. To show that Π is a global homomorphism, consider two elements g, h ∈ SO(3; 1)$^+$. Lay out paths p$_g$, p$_h$ from the identity to them and define a path p$_{gh}$ going along p$_g$(2t) for 0 ≤ t ≤ 1/2 and along p$_g$ · p$_h$(2t - 1) for 1/2 ≤ t ≤ 1. This is a path from the identity to gh. Select adequate partitionings for p$_g$, p$_h$. This corresponds to a choice of "times" t$_0$, t$_1$, ...t$_m$ and s$_0$, s$_1$, ...s$_n$. Divide the first set by 2 and divide the second set by 2 and add 1/2 and so obtain a new (adequate) set of "times" to be used for p$_{gh}$. Direct computation shows that, with these partitionings (and hence all partitionings), Π(gh) = Π(g)Π(h).

##### Surjectiveness of exponential mapping

From a practical point of view, it is important that formula (G2) can be used for all elements of the group. The Lie correspondence theorem above guarantees that (G2) holds for all X ∈ so(3; 1), but provides no guarantee that all g ∈ SO(3; 1)$^+$ are in the image of exp:so(3; 1) → SO(3; 1)$^+$. For general Lie groups, this is not the case, especially not for non-compact groups, as for example for SL(2, C), the universal covering group of SO(3; 1)$^+$. It will be treated in this respect below.
But exp: so(3; 1) → SO(3; 1)$^+$ is surjective. One way to see this is to make use of the isomorphism SO(3; 1)$^+$ ≈ PGL(2, C), the latter being the Möbius group. It is a quotient of GL(n, C) (see the linked article). Let p:GL(n, C) → PGL(2, C) denote the quotient map. Now exp:gl(n, C) → GL(n, C) is onto. Apply the Lie correspondence theorem with π being the differential at the identity of p. Then for all X ∈ gl(n, C) p(e$^{iX}$) = e$^{iπ(X)}$. Since the left hand side is surjective (both exp and p are), the right hand side is surjective and hence exp:pgl(2, C) → PGL(2, C) is surjective. Finally, recycle the argument once more, but now with the known isomorphism between SO(3; 1)$^+$ and PGL(2, C) to find that exp is onto for the connected component of the Lorentz group.

##### Consistency

From the way Π(g) has been defined for elements far from the identity, it not immediately clear that formula (G2) holds for all elements of SO(3; 1)$^+$, i.e. that one can take U = G in (G2). But, in summary,
* Π is a uniquely constructed homomorphism.
* Using (G6) with Π as defined here, then one ends up with the π one started with since Π was defined that way near the identity, and (G6) depends only on an arbitrarily small neighborhood of the identity.
* exp: so(3; 1) → SO(3; 1)$^+$ is surjective.
Hence (G2) holds everywhere. One finally unconditionally writes
Template:Equation box 1

#### Fundamental group

The above construction relies on simple connectedness. The result needs modifications for non-simply connected groups per below. To exhibit the fundamental group of SO(3; 1)$^+$, the topology of its covering group SL(2, C) is considered. By the polar decomposition theorem, any matrix λ ∈ SL(2, C) may be uniquely expressed as
    $\lambda =ue^{h},\det u=1,\operatorname {tr} v=0,$
where u is unitary with determinant one, hence in SU(2), and h is Hermitian with trace zero. The trace and determinant conditions imply
    $h=\left({\begin{matrix}c&a-ib\\a+ibc&-c\end{matrix}}\right),\quad u=\left({\begin{matrix}d+ie&f+ig\\-f+ig&d-ie\end{matrix}}\right),\quad d^{2}+e^{2}+f^{2}+g^{2}=1,$
with (a, b, c) ∈ R$^3$ unconstrained and (d, e, f, g) ∈ R$^4$ constrained to the 3-sphere S$^3$. The manifestly continuous one-to-one map with continuous inverse R$^3$ × S$^3$ → SL(2, C); (r, s) ↦ u(s)e$^{h(r)}$ (a homeomorphism) explicitly exhibits the simple connectedness of SL(2, C). But SO(3; 1) ≈ SL(2, C)/{I, −I}, where {I, −I} is the center of SL(2, C). Identifying λ and −λ amounts to identifying u with −u, which in turn amounts to identifying antipodal points on S$^3$. Thus topologically,
    $SO(3;1)\approx \mathbf {R} ^{3}\times S^{3}/Z_{2},$
where last factor is not simply connected: Geometrically, it is seen (for visualization purposes, replace S$^3$ by S$^2$) that a path from u to −u in SU(2) ≈ S$^3$ is a loop in S$^3$/Z$_2$ since u and −u are antipodal points, and that it is not contractible to a point. But a path from u to −u, thence to u again, a loop in S$^3$ and a double loop (considering p(ue$^h$) = p(−ue$^h$), where p is the covering map SL(2, C) → SL(3; 1)) in S$^3$/Z$_2$ that is contractible to a point (continuously move away from −u "upstairs" in S$^3$ and shrink the path there to the point u). Thus π$_1$(SO(3; 1)) is a two-element group with two equivalence classes of loops as its elements – or put more simply, SO(3; 1) is doubly connected.

#### Projective representations

For a group that is connected but not simply connected, such as SO(3; 1)$^+$, the result may depend on the homotopy class of the chosen path. The result, when using (G2), will then depend on which X in the Lie algebra is used to obtain the representative matrix for g.
Since π$_1$(SO(3; 1)$^+$) per above has two elements, not all representations of the Lie algebra will yield representations of the group, but some will instead yield projective representations. Once these conclusions have been reached, and once it is known whether a representation is projective, there is no need to be concerned about paths and partitions. Formula (G2) applies to all group elements and all representations, including the projective ones.
For the Lorentz group, the (m, n)-representation is projective when m + n is a half-integer. See the section spinors.
For a projective representation Π of SO(3; 1)$^+$, it holds that
Template:Equation box 1
since any loop in SO(3; 1)$^+$ traversed twice, due to the double connectedness, is contractible to a point, so that its homotopy class is that of a constant map. It follows that Π is a double-valued function. It is not possible to consistently chose a sign to obtain a continuous representation of all of SO(3; 1)$^+$, but this is possible locally around any point.

### The covering group

Consider sl(2, C) as a real Lie algebra with basis
    $\left({\frac {1}{\sqrt {2}}}\sigma _{1},{\frac {1}{\sqrt {2}}}\sigma _{2},{\frac {1}{\sqrt {2}}}\sigma _{3},{\frac {i}{\sqrt {2}}}\sigma _{1},{\frac {i}{\sqrt {2}}}\sigma _{2},{\frac {i}{\sqrt {2}}}\sigma _{3}\right)\equiv (j_{1},j_{2},j_{3},k_{1},k_{2},k_{3}),$
where the sigmas are the Pauli matrices. From the relations
    $[\sigma _{i},\sigma _{j}]=2i\epsilon _{ijk}\sigma _{k}$         (J1)
is obtained
    $[j_{i},j_{j}]=i\epsilon _{ijk}j_{k},\quad [j_{i},k_{j}]=i\epsilon _{ijk}k_{k},\quad [k_{i},k_{j}]=-i\epsilon _{ijk}j_{k},$         (J2)
which are exactly on the form of the 3-dimensional version of the commutation relations for so(3; 1) (see conventions and Lie algebra bases below). Thus, one may map J$_i$ ↔ j$_i$, K$_i$ ↔ k$_i$, and extend by linearity to obtain an isomorphism. Since SL(2, C) is simply connected, it is the universal covering group of SO(3; 1)$^+$.

#### A geometric view

Let π$_g$ denote the set of path homotopy classes [p$_g$] of paths p$_g$(t), 0 ≤ t ≤ 1, from 1 ∈ SO(3; 1)$^+$ to g ∈ SO(3; 1)$^+$ and define the set
    $G=\{(g,[p_{g}]):g\in \mathrm {SO} (3;1)^{+},[p_{g}]\in \pi _{g}\}$         (C1)
and endow it with the multiplication operation
    $(g_{1},[p_{1}])(g_{2},[p_{2}])=(g_{1}g_{2},[p_{12}]),\quad g_{1},g_{2}\in \mathrm {SO} (3;1)^{+},\quad [p_{1}]\in \pi _{g_{1}},[p_{2}]\in \pi _{g_{2}},[p_{12}]\in \pi _{g_{12}},\quad p_{12}(t)=p_{1}(t)\cdot p_{2}(t)~.$         (C2)
The dot on the far right denotes path multiplication.
With this multiplication, G is a group and G ≈ SL(2, C), the universal covering group of SO(3; 1)$^+$. By the above construction, there is, since each π$_g$ has two elements, a 2:1 covering map p : G → SO(3; 1)$^+$ and an isomorphism G ≈ SL(2, C). According to covering group theory, the Lie algebras so(3; 1), sl(2, C) and g of G are all isomorphic. The covering map p:G → SO(3; 1)$^+$ is simply given by p(g,[p$_g$]) = g.

#### An algebraic view

For an algebraic view of the universal covering group, let SL(2, C) act on the set of all Hermitian 2×2 matrices h by the operation
    ${\begin{cases}\mathbf {P} (A):\mathbf {h} \to \mathbf {h} \\X\mapsto A^{\dagger }XA\end{cases}}\qquad A\in \mathrm {SL} (2,\mathbf {C} )$         (C3)
Since X ∈ h is Hermitian, A$^†$XA is also Hermitian because (A$^†$XA)$^†$ = A$^†$X$^†$A$^{††}$ = A$^†$XA, and also A$^†$(αX + βY)A = αA$^†$XA + βA$^†$YA, so the action is linear as well. An element of h may generally be written in the form
    $X={\begin{pmatrix}\xi _{4}+\xi _{3}&\xi _{1}+i\xi _{2}\\\xi _{1}-i\xi _{2}&\xi _{4}-\xi _{3}\\\end{pmatrix}}\qquad \xi _{1},\ldots ,\xi _{4}\in \mathbf {R} ,$         (C4)
showing that h is a 4-dimensional real vector space. Moreover, (AB)$^†$X(AB) = B$^†$A$^†$XAB meaning that P is a group homomorphism into GL(h) ⊂ End h. Thus P : SL(2, C) → GL (h) is a 4-dimensional representation of SL(2, C). Its kernel must in particular take the identity matrix to itself, A$^†$IA = A$^†$A = I and therefore A$^†$ = A$^{−1}$. Thus AX = XA for A in the kernel so, by Schur's lemma, A is a multiple of the identity, which must be ±I since det A = 1. Now map h to spacetime R$^4$ endowed with the Lorentz metric, Minkowski space, via
    $X=(\xi _{1},\xi _{2},\xi _{3},\xi _{4})\leftrightarrow {\overrightarrow {(\xi _{1},\xi _{2},\xi _{3},\xi _{4})}}=(x,y,z,t)={\overrightarrow {X}}.$         (C5)
The action of P(A) on h preserves determinants since det(A$^†$XA) = (det A)(det A$^†$)(det X) = det X. The induced representation p of SL(2, C) on R$^4$, via the above isomorphism, given by
    $\mathbf {p} (A){\overrightarrow {X}}={\overrightarrow {AXA^{\dagger }}}$         (C6)
will preserve the Lorentz inner product since
    $-\det X=\xi _{1}^{2}+\xi _{2}^{2}+\xi _{3}^{2}-\xi _{4}^{2}=x^{2}+y^{2}+z^{2}-t^{2}.$
This means that p(A) belongs to the full Lorentz group SO(3; 1). By the main theorem of connectedness, since SL(2, C) is connected, its image under p in SO(3; 1) is connected as well, and hence is contained in SO(3; 1)$^+$.
It can be shown that the Lie map of p : SL(2, C) → SO(3; 1)$^+$, π : sl(2, C) → so(3; 1) is a Lie algebra isomorphism The map P is also onto.
Thus SL(2, C), since it is simply connected, is the universal covering group of SO(3; 1)$^+$, isomorphic to the group G of above.

#### Representations of SL(2, C) and sl(2, C)

The complex linear representations of sl(2, C) and SL(2, C) are more straightforward to obtain than the so(3; 1)$^+$ representations. They can be (and usually are) written down from scratch. The holomorphic group representations (meaning the corresponding Lie algebra representation is complex linear) are related to the complex linear Lie algebra representations by exponentiation. The real linear representations of sl(2, C) are exactly the (μ, ν)-representations. They can be exponentiated too. The (μ, 0)-representations are complex linear and are (isomorphic to) the highest weight-representations. These are usually indexed with only one integer (but half-integers are used here).
The mathematics convention is used in this section for convenience. Lie algebra elements differ by a factor of i and there is no factor of i in the exponential mapping compared to the physics convention used elsewhere. Let the basis of sl(2, C) be
    $H={\bigl{\begin{smallmatrix}1&0\\0&-1\\\end{smallmatrix}}{\bigr )},\quad X={\bigl{\begin{smallmatrix}0&1\\0&0\\\end{smallmatrix}}{\bigr )},\quad Y={\bigl{\begin{smallmatrix}0&0\\1&0\\\end{smallmatrix}}{\bigr )}.$         (S1)
This choice of basis, and the notation, is standard in the mathematical literature.

##### Complex linear representations

The irreducible holomorphic (n + 1)-dimensional representations of SL(2, C), n ≥ 0, can be realized on the space of homogeneous polynomial of degree n in 2 variables ℙ$^2$$_n$, the elements of which appear as
    $P(z_{1},z_{2})=c_{n}z_{1}^{n}+c_{n-1}z_{1}^{n-1}z_{2}+\ldots +c_{1}z_{2}^{n}.$
The action of SL(2, C) is given by
    $[\Phi _{n}(g)]P(z)={\biggl\Phi {\biggl{\begin{matrix}a&b\\c&d\\\end{matrix}}{\biggr )}P{\biggr ]}{\biggl{\begin{matrix}z_{1}\\z_{2}\\\end{matrix}}{\biggr )}=P{\biggl{\biggl{\begin{matrix}a&b\\c&d\\\end{matrix}}{\biggr )}^{-1}{\biggl{\begin{matrix}z_{1}\\z_{2}\\\end{matrix}}{\biggr )}{\biggr )},\quad P\in \mathbf {P} _{n}^{2}.$         (S2)
The associated sl(2, C)-action is, using (G6) and the definition above, for the basis elements of sl(2, C),
    $\phi _{n}(H)=-z_{1}{\frac {\partial }{\partial z_{1}}}+z_{2}{\frac {\partial }{\partial z_{2}}},\quad \phi _{n}(X)=-z_{2}{\frac {\partial }{\partial z_{1}}},\quad \phi _{n}(Y)=-z_{1}{\frac {\partial }{\partial z_{2}}}.$         (S5)
With a choice of basis for P ∈ ℙ$^2$$_n$, these representations become matrix Lie algebras.

##### Real linear representations

The (μ, ν)-representations are realized on a space of polynomials ℙ$^2$$_{μν}$ in z$_1$, z$_1$, z$_2$, z$_2$, homogeneous of degree μ in z$_1$, z$_2$ and homogeneous of degree ν in z$_1$, z$_2$. The representations are given by
Template:Equation box 1
By employing (G6) again it is found that
Template:Equation box 1
from which the expressions
Template:Equation box 1
for the basis elements follow.

##### Non-surjectiveness of exponential mapping

Unlike in the case exp: so(3; 1) → SO(3; 1)$^+$, the exponential mapping exp: sl(2, C) → SL(2, C) is not onto. The conjugacy classes of SL(2, C) are represented by the matrices
    $a(\alpha )={\begin{pmatrix}\alpha &0\\0&\alpha ^{-1}\\\end{pmatrix}}=e^{\log(\alpha )H},\alpha \neq 0,\quad p={\begin{pmatrix}1&1\\0&1\\\end{pmatrix}}=e^{X},\quad q={\begin{pmatrix}-1&1\\0&-1\\\end{pmatrix}},$         (S6)
but there is no element Q in sl(2, C) such that q = exp(Q).
In general, if g is an element of a connected Lie group G with Lie algebra g, then
    $g=e^{X_{1}}e^{X_{2}}\cdots e^{X_{n}},\quad X_{i}\in {\mathfrak {g}},1\leq i\leq n.$         (S7)
This follows from the compactness of a path from the identity to g and the one-to-one nature of exp near the identity. In the case of the matrix q, one may write
    $e^{-X}e^{i\pi H}=e^{{\bigl{\begin{smallmatrix}0&-1\\0&0\\\end{smallmatrix}}{\bigr )}}e^{i\pi {\bigl{\begin{smallmatrix}1&0\\0&-1\\\end{smallmatrix}}{\bigr )}}={\begin{pmatrix}1&-1\\0&1\\\end{pmatrix}}{\begin{pmatrix}-1&0\\0&-1\\\end{pmatrix}}={\begin{pmatrix}-1&1\\0&-1\\\end{pmatrix}}=q.$         (S8)
The kernel of the covering map p:SL(2, C) → SO(3; 1)$^+$ of above is N = {I, −I}, a normal subgroup of SL(2, C)$^+$. The composition p ∘ exp: sl(2, C) → SO(3; 1) is onto. If a matrix a is not in the image of exp, then there is a matrix b equivalent to it with respect to p, meaning p(b) = p(a), that is in the image of exp. The condition for equivalence is a$^{−1}$b ∈ N. In the case of the matrix q, one may solve for p in the equation p$^{−1}$q = −I ∈ N. One finds
    $p={\begin{pmatrix}1&-1\\0&1\\\end{pmatrix}}=e^{-X},\quad p^{-1}q={\begin{pmatrix}1&1\\0&1\\\end{pmatrix}}{\begin{pmatrix}-1&1\\0&-1\\\end{pmatrix}}={\begin{pmatrix}-1&0\\0&-1\\\end{pmatrix}}=-I.$         (S9)
As a corollary, since the covering map p is a homomorphism,the mapping version of the Lie correspondence (G6) can be used to provide a proof of the surjectiveness of exp for so(3; 1). Let σ denote the isomorphism between sl(2, C) and so(3; 1). Refer to the commutative diagram. One has p ∘ exp: sl(2, C) → SO(3; 1) = exp ∘ σ for all X ∈ sl(2, C). Since p ∘ exp is onto, exp ∘ σ is onto, and hence exp: so(3; 1) → SO(3; 1)$^+$ is onto as well.

##### SO(3; 1)$^+$-representations from SL(2, C)-representations

By the first isomorphism theorem, a representation (Φ, V) of SL(2, C) descends to a representation (Π, V) of SO(3; 1)$^+$ if and only if ker p ⊂ ker Φ. Refer to the commutative diagram. If this condition holds, then both elements in the fiber p$^{−1}$(g), g ∈ SO(3; 1)$^+$ will be mapped by Φ to the same representative, and the expression Φ(p$^{−1}$(g)) makes sense. One may thus define Π: SO(3; 1)$^+$ → GL(V), Π(g) = Φ(p$^{−1}$(g)). In particular, if Π is faithful, i.e. having kernel = I, then there is no corresponding proper representation of SO(3; 1)$^+$, but there is a projective one as was shown in a previous section, corresponding to the two possible choices of representative in each fiber p$^{−1}$(g).
Lie algebra representations of so(3; 1) are obtained from sl(2, C)-representations simply by composition with σ$^{−1}$.

##### SL(2, C)-representations from SO(3; 1)$^+$-representations

SL(2, C)-representations can be obtained from non-projective SO(3; 1)$^+$-representations by composition with the projection map p. These are always representations since they are compositions of group homomorphisms. Such a representation is never faithful because Ker p = {I, −I}. If the SO(3; 1)$^+$-representation is projective, then the resulting SL(2, C)-representation would be projective as well. Instead, the isomorphism σ:so(3; 1) → sl(3, C) can be employed, composed with exp:sl(2, C) → SL(2, C). This is always a non-projective representation.

### Properties of the (m, n) representations

The (m, n) representations are irreducible, and they are the only irreducible representations.
* Irreducibility follows from the unitarian trick and that a representation Π of SU(2) × SU(2) is irreducible if and only if Π = Π$_μ$ ⊗ Π$_ν$, where Π$_μ$, Π$_ν$ are irreducible representations of SU(2).
* Uniqueness follows from that the Π$_m$ are the only irreducible representations of SU(2), which is one of the conclusions of the theorem of the highest weight.

#### Dimension

The (m, n) representations are (2m + 1)(2n + 1)-dimensional. It follows from the Weyl dimension formula. For a Lie algebra g it reads
    $\dim \pi _{\mu }={\frac {\Pi _{\alpha \in R^{+}}\langle \alpha ,\mu +\delta \rangle }{\Pi _{\alpha \in R^{+}}\langle \alpha ,\delta \rangle }},$
where R$^+$ is the set of positive roots and δ is half the sum of the positive roots. The inner product $\langle \cdot ,\cdot \rangle$ is that of the Lie algebra g, invariant under the action of the Weyl group on h ⊂ g, the Cartan subalgebra. The roots (really elements of h*) are via this inner product identified with elements of h. For sl(2, C), the formula reduces to dim π$_μ$ = μ + 1 = 2m + 1. By taking tensor products, the result follows.
A quicker approach is, of course, to simply count the dimensions in any concrete realization, such as the one given in representations of SL(2, C) and sl(2,  C).

#### Faithfulness

If a representation Π of a Lie group G is not faithful, then N = ker Π is a nontrivial normal subgroup because Π(n) = I ⇒ Π(gng$^{−1}$) = Π(g)Π(n)Π(g)$^{−1}$ = Π(g)Π(g)$^{−1}$ = I. There are three relevant cases.
* N is non-discrete and abelian.
* N is non-discrete and non-abelian.
* N is discrete. In this case N ⊂ Z, where Z is the center of G.
In the case of SO(3; 1)$^+$, the first case is excluded since SO(3; 1)$^+$ is semi-simple. The second case (and the first case) is excluded because SO(3; 1)$^+$ is simple. For the third case, SO(3; 1)$^+$ is isomorphic to the quotient SL(2, C)/{I, −I}. But {I, −I} is the center of SL(2, C). It follows that the center of SO(3; 1)$^+$ is trivial, and this excludes the third case. The conclusion is that every representation Π:SO(3; 1)$^+$ → GL(V) and every projective representation Π:SO(3; 1)$^+$ → PGL(W) for V, W finite-dimensional vector spaces are faithful.
By using the fundamental Lie correspondence, the statements and the reasoning above translate directly to Lie algebras with (abelian) nontrivial non-discrete normal subgroups replaced by (one-dimensional) nontrivial ideals in the Lie algebra, and the center of SO(3; 1)$^+$ replaced by the center of sl(3; 1)$^+$. The center of any semisimple Lie algebra is trivial and so(3; 1) is semi-simple and simple, and hence has no non-trivial ideals.
A related fact is that if the corresponding representation of SL(2, C) is faithful, then the representation is projective. Conversely, if the representation is non-projective, then the corresponding SL(2, C) representation is not faithful, but is 2:1.

#### Non-unitarity

The (m, n) Lie algebra representation is not Hermitian. Accordingly, the corresponding (projective) representation of the group is never unitary. This is due to the non-compactness of the Lorentz group. In fact, a connected simple non-compact Lie group cannot have any nontrivial unitary finite-dimensional representations. There is a topological proof of this. Let U:G → GL(V), where V is finite-dimensional, be a continuous unitary representation of the non-compact connected simple Lie group G. Then U(G) ⊂ U(V) ⊂ GL(V) where U(V) is the compact subgroup of GL(V) consisting of unitary transformations of V. The kernel, ker U, of U is a normal subgroup of G. Since G is simple, ker U is either all of G, in which case U is trivial, or ker U is trivial, in which case U is faithful. In the latter case U is a diffeomorphism onto its image, U(G) ≈ G., and U(G) is Lie group. This would mean that U(G) is an embedded non-compact Lie subgroup of the compact group U(V). This is impossible with the subspace topology on U(G) ⊂ U(V) since all embedded Lie subgroups of a Lie group are closed If U(G) were closed, it would be compact, and then G would be compact, contrary to assumption.
In the case of the Lorentz group, this can also be seen directly from the definitions. The representations of A and B used in the construction are Hermitian. This means that J is Hermitian, but K is anti-Hermitian. The non-unitarity is not a problem in quantum field theory, since the objects of concern are not required to have a Lorentz-invariant positive definite norm.

#### Restriction to SO(3)

The (m, n) representation is, however, unitary when restricted to the rotation subgroup SO(3), but these representations are not irreducible as representations of SO(3). A Clebsch–Gordan decomposition can be applied showing that an (m, n) representation have SO(3)-invariant subspaces of highest weight (spin) m + n, m + n − 1, … , | m − n |, where each possible highest weight (spin) occurs exactly once. A weight subspace of highest weight (spin) j is (2j + 1)-dimensional. So for example, the (1/2, 1/2) representation has spin 1 and spin 0 subspaces of dimension 3 and 1 respectively.
Since the angular momentum operator is given by J = A + B, the highest spin in quantum mechanics of the rotation sub-representation will be (m + n)ℏ and the "usual" rules of addition of angular momenta and the formalism of 3-j symbols, 6-j symbols, etc. applies.

#### Spinors

It is the SO(3)-invariant subspaces of the irreducible representations that determine whether a representation has spin. From the above paragraph, it is seen that the (m, n) representation has spin if m + n is half-integral. The simplest are ( 1/2, 0) and (0, 1/2), the Weyl-spinors of dimension 2. Then, for example, (0, 3/2) and (1, 1/2) are a spin representations of dimensions 23/2 + 1 = 4 and (2 + 1)(21/2 + 1) = 6 respectively. Note that, according to the above paragraph, there are subspaces with spin both 3/2 and 1/2 in the last two cases, so these representations cannot likely represent a single physical particle which must be well-behaved under SO(3). It cannot be ruled out in general, however, that representations with multiple SO(3) subrepresentations with different spin can represent physical particles with well-defined spin. It may be that there is a suitable relativistic wave equation that projects out unphysical components, leaving only a single spin.
Construction of pure spin n/2 representations for any n (under SO(3)) from the irreducible representations involves taking tensor products of the Dirac-representation with a non-spin representation, extraction of a suitable subspace, and finally imposing differential constraints.

#### Dual representations

To see if the dual representation of an irreducible representation is isomorphic to the original representation one can consider the following theorems:
* The set of weights of the dual representation of an irreducible representation of a semisimple Lie algebra is, including multiplicities, the negative of the set of weights for the original representation.
* Two irreducible representations are isomorphic if and only if they have the same highest weight.
* For each semisimple Lie algebra there exists a unique element w$_0$ of the Weyl group such that if μ is a dominant integral weight, then w$_0$ ⋅ (−μ) is again a dominant integral weight.
* If π$_{μ0}$ is an irreducible representation with highest weight μ$_0$, then π*$_{μ0}$ has highest weight w$_0$ ⋅ (−μ).
Here, the elements of the Weyl group are considered as orthogonal transformations, acting by matrix multiplication, on the real vector space of roots. One sees that if −I is an element of the Weyl group of a semisimple Lie algebra, then w$_0$ = −I. In the case of sl(2, C), the Weyl group is W = {I, −I}. It follows that each π$_μ$, μ = 0, 1, … is isomorphic to its dual π$_μ$*. The root system of sl(2, C) ⊕ sl(2, C) is shown in the figure to the right. The Weyl group is generated by {w$_γ$} where w$_γ$ is reflection in the plane orthogonal to γ as γ ranges over all roots. One sees that w$_α$ ⋅w$_β$ = −I so −I ∈ W. Then using the fact that if π, σ are Lie algebra representations and π ≈ σ, then Π ≈ Σ. The conclusion for SO(3; 1)$^+$ is
    $\pi _{m,n}^{*}\cong \pi _{m,n},\quad \Pi _{m,n}^{*}\cong \Pi _{m,n},\quad 2m,2n\in \mathbf {N} .$

#### Complex conjugate representations

If π is a representation of a Lie algebra, then π is a representation, where the bar denotes entry-wise complex conjugation in the representative matrices. This follows from that complex conjugation commutes with addition and multiplication. In general, every irreducible representation π of sl(n, C) can be written uniquely as π = π$^+$ + π$^−$, where
    $\pi ^{\pm }(X)={\frac {1}{2}}(\pi (X)\pm i\pi (i^{-1}X)),$
with π$^+$ holomorphic (complex linear) and π$^−$ anti-holomorphic (conjugate linear). For sl(2, C), since π$_μ$ is holomorphic, π$_μ$ is anti-holomorphic. Direct examination of the explicit expressions for π$_{μ,0}$ and π$_{0,ν}$ in equation (S8) below shows that they are holomorphic and anti-holomorphic respectively. Closer examination of the expression (S8) also allows for identification of π$^+$ and π$^−$ for π$_{μ,ν}$ as π$^+$$_{μ,ν}$ = π$_μ$$^{⊕ν+1}$ and π$^−$$_{μ,ν}$ = π$_ν$$^{⊕μ+1}$.
Using the above identities (interpreted as pointwise addition of functions), for SO(3; 1)$^+$ yields
    ${\overline {\pi _{m,n}}}={\overline {\pi _{m,n}^{+}+\pi _{m,n}^{-}}}={\overline {\pi _{m}^{\oplus _{2n+1}}}}+{\overline {{\overline {\pi _{n}}}^{\oplus _{2m+1}}}}=\pi _{n}^{\oplus _{2m+1}}+{\overline {\pi _{m}}}^{\oplus _{2n+1}}=\pi _{n,m}^{+}+\pi _{n,m}^{-}=\pi _{n,m},\quad {\overline {\Pi _{m,n}}}=\Pi _{n,m},\quad 2m,2n\in \mathbf {N} ,$
where the statement for the group representations follow from exp(X) = exp(X). It follows that the irreducible representations (m, n) have real matrix representatives if and only if m = n. Reducible representations on the form (m, n) ⊕ (n, m) have real matrices too.

### The adjoint representation, the Clifford algebra, and the Dirac spinor representation

In general representation theory, if (π, V) is a representation of a Lie algebra g, then there is an associated representation of g on End V, also denoted π, given by
    $\pi (X)(A)=[\pi (X),A],\quad A\in \operatorname {End} V,\ X\in {\mathfrak {g}}.$         (I1)
Likewise, a representation (Π, V) of a group G yields a representation Π on End V of G, still denoted Π, given by
    $\Pi (g)(A)=\Pi (g)A\Pi (g)^{-1},\quad A\in \operatorname {End} V,\ g\in G.$         (I2)
If π and Π are the standard representations on R$^4$ and if the action is restricted to so(3, 1) ⊂ End R$^4$, then the two above representations are the adjoint representation of the Lie algebra and the adjoint representation of the group respectively. The corresponding representations (some R$^n$ or C$^n$) always exist for any matrix Lie group, and are paramount for investigation of the representation theory in general, and for any given Lie group in particular.
Applying this to the Lorentz group, if (Π, V) is a projective representation, then direct calculation using (G4) shows that the induced representation on End V is, in fact, a proper representation, i.e. a representation without phase factors.
In quantum mechanics this means that if (π, H) or (Π, H) is a representation acting on some Hilbert space H, then the corresponding induced representation acts on the set of linear operators on H. As an example, the induced representation of the projective spin (1/2, 0) ⊕ (0, 1/2) representation on End(H) is the non-projective 4-vector (1/2, 1/2) representation.
For simplicity, consider now only the "discrete part" of End H, that is, given a basis for H, the set of constant matrices of various dimension, including possibly infinite dimensions. A general element of the full End H is the sum of tensor products of a matrix from the simplified End H and an operator from the left out part. The left out part consists of functions of spacetime, differential and integral operators and the like. See Dirac operator for an illustrative example. Also left out are operators corresponding to other degrees of freedom not related to spacetime, such as gauge degrees of freedom in gauge theories.
The induced 4-vector representation of above on this simplified End H has an invariant 4-dimensional subspace that is spanned by the four gamma matrices. (Note the different metric convention in the linked article.) In a corresponding way, the complete Clifford algebra of spacetime, Cℓ$_{3,1}$(R), whose complexification is M$_4$(C), generated by the gamma matrices decomposes as a direct sum of representation spaces of a scalar irreducible representation (irrep), the (0, 0), a pseudoscalar irrep, also the (0, 0), but with parity inversion eigenvalue −1, see the next section below, the already mentioned vector irrep, (1/2, ,1/2), a pseudovector irrep, (1/2, 1/2) with parity inversion eigenvalue +1 (not −1), and a tensor irrep, (1, 0) ⊕ (0, 1). The dimensions add up to 1 + 1 + 4 + 4 + 6 = 16. In other words,
    $Cl_{3,1}(\mathbf {R} )=(0,0)\oplus ({\frac {1}{2}},{\frac {1}{2}})\oplus [(1,0)\oplus (0,1)]\oplus ({\frac {1}{2}},{\frac {1}{2}})_{p}\oplus (0,0)_{p},$         (I3)
where, as is customary, a representation is confused with its representation space. This is, in fact, a reasonably convenient way to show that the algebra spanned by the gammas is 16-dimensional.

#### The (1/2, 0) ⊕ (0, 1/2) spin representation

The six-dimensional representation space of the tensor (1, 0) ⊕ (0, 1)-representation inside Cℓ$_{3,1}$(R) has two roles. In particular, letting
    $\sigma ^{\mu \nu }=-{\frac {i}{4}}[\gamma ^{\mu },\gamma ^{\nu }],$         (I4)
where {γ$^μ$ ∈ Cℓ$_{3,1}$(R): μ = 0,1,2,3} are the gamma matrices, the {σ$^{μν}$ ∈ Cℓ$_{3,1}$(R)} , only 6 of which are non-zero due to antisymmetry of the bracket, span the tensor representation space. Moreover, they have the commutation relations of the Lorentz Lie algebra,
    $[\sigma ^{\mu \nu },\sigma ^{\rho \tau }]=i(\eta ^{\tau \mu }\sigma ^{\rho \nu }+\eta ^{\nu \tau }\sigma ^{\mu \rho }-\eta ^{\rho \mu }\sigma ^{\tau \nu }-\eta ^{\nu \rho }\sigma ^{\mu \tau }),$         (I5)
and hence constitute a representation (in addition to being a representation space) sitting inside Cℓ$_{3,1}$(R), the (1/2, 0) ⊕ (0, 1/2) spin representation. For details, see bispinor and Dirac algebra.
The conclusion is that every element of the complexified Cℓ$_{3,1}$(R) in End H (i.e. every complex 4×4 matrix) has well defined Lorentz transformation properties. In addition, it has a spin-representation of the Lorentz Lie algebra, which upon exponentiation becomes a spin representation of the group, acting on C$^4$, making it a space of bispinors.

### Reducible representations

There is a multitude of other representations that can be deduced from the irreducible ones, such as those obtained in a standard manner by taking direct sums, tensor products, and quotients of the irreducible representations. Other methods of obtaining representations include the restriction of a representation of a larger group containing the Lorentz group, e.g. GL(n, R). These representations are in general not irreducible, and are not discussed here. It is to be noted though that the Lorenz group and its Lie algebra have the complete reducibility property. This means that every representation reduces to a direct sum of irreducible representations.

### Space inversion and time reversal

The (possibly projective) (m, n) representation is irreducible as a representation SO(3; 1)$^+$, the identity component of the Lorentz group, in physics terminology the proper orthochronous Lorentz group. If m = n it can be extended to a representation of all of O(3; 1), the full Lorentz group, including space parity inversion and time reversal. The representations (m, n) ⊕ (n, m) can be extended likewise.

#### Space parity inversion

For space parity inversion, one considers the adjoint action Ad$_P$ of P ∈ SO(3; 1) on so(3; 1), where P is the standard representative of space parity inversion, P = diag(1, −1, −1, −1), given by
    $\mathrm {Ad} _{P}(J_{i})=PJ_{i}P^{-1}=J_{i},\qquad \mathrm {Ad} _{P}(K_{i})=PK_{i}P^{-1}=-K_{i}.$         (F1)
It is these properties of K and J under P that motivate the terms vector for K and pseudovector or axial vector for J. In a similar way, if π is any representation of so(3; 1) and Π is its associated group representation, then Π(SO(3; 1)$^+$) acts on the representation of π by the adjoint action, π(X) ↦ Π(g) π(X) Π(g)$^{−1}$ for X ∈ so(3; 1), g ∈ SO(3; 1)$^+$. If P is to be included in Π, then consistency with (F1) requires that
    $\Pi (P)\pi (B_{i})\Pi (P)^{-1}=\pi (A_{i})$         (F2)
holds, where A and B are defined as in the first section. This can hold only if A$_i$ and B$_i$ have the same dimensions, i.e. only if m = n. When m ≠ n then (m, n) ⊕ (n, m) can be extended to an irreducible representation of SO(3; 1)$^+$, the orthocronous Lorentz group. The parity reversal representative Π(P) does not come automatically with the general construction of the (m, n) representations. It must be specified separately. The matrix β = i γ$^0$ (or a multiple of modulus −1 times it) may be used in the (1/2, 0) ⊕ (0, 1/2) representation.
If parity is included with a minus sign (the 1×1 matrix [−1]) in the (0,0) representation, it is called a pseudoscalar representation.

#### Time reversal

Time reversal T = diag(−1, 1, 1, 1), acts similarly on so(3; 1) by
    $\mathrm {Ad} _{T}(J_{i})=TJ_{i}T^{-1}=-J_{i},\qquad \mathrm {Ad} _{T}(K_{i})=TK_{i}T^{-1}=K_{i}.$         (F3)
By explicitly including a representative for T, as well as one for P, one obtains a representation of the full Lorentz group O(3; 1). A subtle problem appears however in application to physics, in particular quantum mechanics. When considering the full Poincaré group, four more generators, the P$^μ$, in addition to the J$^i$ and K$^i$ generate the group. These are interpreted as generators of translations. The time-component P$^0$ is the Hamiltonian H. The operator T satisfies the relation
    $\mathrm {Ad} _{T}(iH)=TiHT^{-1}=-iH$         (F4)
in analogy to the relations above with so(3; 1) replaced by the full Poincaré algebra. By just cancelling the i's, the result THT$^{−1}$ = −H would imply that for every state Ψ with positive energy E in a Hilbert space of quantum states with time-reversal invariance, there would be a state Π(T$^{−1}$)Ψ with negative energy −E. Such states do not exist. The operator Π(T) is therefore chosen antilinear and antiunitary, so that it anticommutes with i, resulting in THT$^{−1}$ = +H, and its action on Hilbert space likewise becomes antilinear and antiunitary. It may be expressed as the composition of complex conjugation with multiplication by a unitary matrix. This is mathematically sound, see Wigner's theorem, but if one is very strict with terminology, Π is not a representation.
When constructing theories such as QED which is invariant under space parity and time reversal, Dirac spinors may be used, while theories that do not, such as the electroweak force, must be formulated in terms of Weyl spinors. The Dirac representation, (1/2, 0) ⊕ (0, 1/2), is usually taken to include both space parity and time inversions. Without space parity inversion, it is not an irreducible representation.
The third discrete symmetry entering in the CPT theorem along with P and T, charge conjugation symmetry C, has nothing directly to do with Lorentz invariance.

## Action on function spaces

In the classification of the irreducible finite-dimensional representations of above it was never specified precisely how a representative of a group or Lie algebra element acts on vectors in the representation space. The action can be anything as long as it is linear. The point silently adopted was that after a choice of basis in the representation space, everything becomes matrices anyway.
If V is a vector space of functions of a finite number of variables n, then the action on a scalar function f ∈ V given by
    $(\Pi (g)f)(x)=f(\Pi _{x}(g)^{-1}x),\qquad x\in \mathbf {R} ^{n},f\in V$         (H1)
produces another function Πf ∈ V. Here Π$_x$ is an n-dimensional representation, and Π is a possibly infinite-dimensional representation. A special case of this construction is when V is a space of functions defined on the group G itself, viewed as a n-dimensional manifold embedded in R$^n$. This is the setting in which the Peter–Weyl theorem and the Borel–Weil theorem are formulated. The former demonstrates the existence of a Fourier decomposition of functions on a compact group into characters of finite-dimensional representations. The completeness of the characters in this sense can thus be used to prove the existence of the highest weight representations. The latter theorem, providing more explicit representations, makes use of the unitarian trick to yield representations of complex non-compact groups, e.g. SL(2, C); in the present case, there is a one-to-one correspondence between representations of SU(2) and holomorphic representations of SL(2, C). (A group representation is called holomorphic if its corresponding Lie algebra representation is complex linear.) This theorem too can be used to demonstrate the existence of the highest weight representations.

### Euclidean rotations

The subgroup SO(3) of three-dimensional Euclidean rotations has an infinite-dimensional representation on the Hilbert space
    $L^{2}(\mathbf {S} ^{2})={\text{span}}\left\{Y_{m}^{l},l\in \mathbf {N} ^{+},-l\leqslant m\leqslant l\right\},$
where $Y_{m}^{l}$ are spherical harmonics. Its elements are square integrable complex-valued functions on the sphere. The inner product on this space is given by
    $\langle f,g\rangle =\int _{\mathbf {S} ^{2}}{\overline {f}}gd\Omega =\int _{0}^{2\pi }\int _{0}^{\pi }{\overline {f}}g\sin \theta d\theta d\varphi .$         (H1)
If f is an arbitrary square integrable function defined on the unit sphere S$^2$, then it can be expressed as
    $|f\rangle =\sum _{l=1}^{\infty }\sum _{m=-l}^{m=l}|Y_{m}^{l}\rangle \langle Y_{m}^{l}|f\rangle ,\qquad f(\theta ,\varphi )=\sum _{l=1}^{\infty }\sum _{m=-l}^{m=l}f_{lm}Y_{m}^{l}(\theta ,\varphi ),$         (H2)
where the expansion coefficients are given by
    $f_{lm}=\langle Y_{m}^{l},f\rangle =\int _{\mathbf {S} ^{2}}{\overline {Y_{m}^{l}}}fd\Omega =\int _{0}^{2\pi }\int _{0}^{\pi }{\overline {Y_{m}^{l}}}(\theta ,\varphi )f(\theta ,\varphi )\sin \theta d\theta d\varphi .$         (H3)
The Lorentz group action restricts to that of SO(3) and is expressed as
    $(\Pi (R)f)(\theta (x),\varphi (x))=\sum _{l=1}^{\infty }\sum _{m=-l}^{m=l}\sum _{m'=-l}^{m'=l}D_{mm'}^{(l)}(R)f_{lm'}Y_{m}^{l}\left(\theta (R^{-1}x),\varphi (R^{-1}x)\right),\qquad R\in \mathrm {SO} (3),\quad x\in \mathbf {S} ^{2}.$         (H4)
This action is unitary, meaning that
    $\langle \Pi (R)f,\Pi (R)g\rangle =\langle f,g\rangle \qquad \forall f,g\in \mathbf {S} ^{2},\quad \forall R\in \mathrm {SO} (3).$         (H5)
The D can be obtained from the D$^{(m,n)}$ of above using Clebsch–Gordan decomposition, but they are more easily directly expressed as an exponential of an odd-dimensional su(2)-representation (the 3-dimensional one is exactly so(3)). In this case the space L$^2$(S$^2$) decomposes neatly into an infinite direct sum of irreducible odd finite-dimensional representations V$_{2i+1}$, i = 0, 1, … according to
    $L^{2}(\mathbf {S} ^{2})=\sum _{i=0}^{\infty }V_{2i+1}\equiv \bigoplus _{i=0}^{\infty }\operatorname {span} \left\{Y_{m}^{2i+1}\right\}.$         (H6)
This is characteristic of infinite-dimensional unitary representations of SO(3). If Π is an infinite-dimensional unitary representation on a separable Hilbert space, then it decomposes as a direct sum of finite-dimensional unitary representations. Such a representation is thus never irreducible. All irreducible finite-dimensional representations (Π, V) can be made unitary by an appropriate choice of inner product,
    $\langle f,g\rangle _{U}\equiv \int _{\mathrm {SO} (3)}\langle \Pi (R)f,\Pi (R)g\rangle dg={\frac {1}{8\pi ^{2}}}\int _{0}^{2\pi }\int _{0}^{\pi }\int _{0}^{2\pi }\langle \Pi (R)f,\Pi (R)g\rangle \sin \theta d\varphi d\theta d\psi ,\quad f,g\in V,$
where the integral is the unique invariant integral over SO(3) normalized to 1, here expressed using the Euler angles parametrization. The inner product inside the integral is any inner product on V.

### The Möbius group

The identity component of the Lorentz group is isomorphic to the Möbius group M. This group can be thought of as conformal mappings of either the complex plane or, via stereographic projection, the Riemann sphere. In this way, the Lorentz group itself can be thought of as acting conformally on the complex plane or on the Riemann sphere.
In the plane, a Möbius transformation characterized by the complex numbers a, b, c, d acts on the plane according to
    $f(z)={\frac {az+b}{cz+d}},\qquad ad-bc\neq 0$.         (M1)
and can be represented by complex matrices
    $\Pi _{f}={\begin{pmatrix}A&B\\C&D\end{pmatrix}}=\lambda {\begin{pmatrix}a&b\\c&d\end{pmatrix}},\qquad \lambda \in \mathbf {C} -\{0\},\,\operatorname {det} \Pi _{f}=1,$         (M2)
since multiplication by a nonzero complex scalar does not change f. These are elements of SL(2,  C) and are unique up to a sign (since ±Π$_f$ give the same f), hence M ≈ SL(2,  C)/{I,  −I} ≈ SO(3; 1)$^+$.

### The Riemann P-functions

The Riemann P-functions, solutions of Riemann's differential equation, are an example of a set of functions that transform among themselves under the action of the Lorentz group. The Riemann P-functions are expressed as
    $w(z)=P\left\{{\begin{matrix}a&b&c&\;\\\alpha &\beta &\gamma &z\\\alpha '&\beta '&\gamma '&\;\end{matrix}}\right\}=\left({\frac {z-a}{z-b}}\right)^{\alpha }\left({\frac {z-c}{z-b}}\right)^{\gamma }P\left\{{\begin{matrix}0&\infty &1&\;\\0&\alpha +\beta +\gamma &0&\;{\frac {(z-a)(c-b)}{(z-b)(c-a)}}\\\alpha '-\alpha &\alpha +\beta '+\gamma &\gamma '-\gamma &\;\end{matrix}}\right\},$         (T1)
where the a,  b,  c,  α,  β,  γ,  α′,  β′,  γ′ are complex constants. The P-function on the right hand side can be expressed using standard hypergeometric functions. The connection is
    $P\left\{{\begin{matrix}0&\infty &1&\;\\0&a&0&z\\1-c&b&c-a-b&\;\end{matrix}}\right\}=\;_{2}F_{1}(a,b;c;z).$         (T2)
The set of constants 0, ∞, 1 in the upper row on the left hand side are the regular singular points of the Gauss' hypergeometric equation. Its exponents, i. e. solutions of the indicial equation, for expansion around the singular point 0 are 0 and 1 − c ,corresponding to the two linearly independent solutions, and for expansion around the singular point 1 they are 0 and c − a − b. Similarly, the exponents for ∞ are a and b for the two solutions.
One has thus
    $w(z)=\left({\frac {z-a}{z-b}}\right)^{\alpha }\left({\frac {z-c}{z-b}}\right)^{\gamma }\;_{2}F_{1}\left(\alpha +\beta +\gamma ,\alpha +\beta '+\gamma ;1+\alpha -\alpha ';{\frac {(z-a)(c-b)}{(z-b)(c-a)}}\right),$         (T3)
where the condition (sometimes called Riemann's identity)
    $\alpha +\alpha '+\beta +\beta '+\gamma +\gamma '=1$
on the exponents of the solutions of Riemann's differential equation has been used to define γ′.
The first set of constants on the left hand side in (T1), a, b, c denotes the regular singular points of Riemann's differential equation. The second set, α, β, γ, are the corresponding exponents at a, b, c for one of the two linearly independent solutions, and, accordingly, α′, β′, γ′ are exponents at a, b, c for the second solution.
Define an action of the Lorentz group on the set of all Riemann P-functions by first setting
    $u(\Lambda )(z)={\frac {Az+B}{Cz+D}},$         (T4)
where A,  B,  C,  D are the entries in
    $\lambda ={\begin{pmatrix}A&B\\C&D\end{pmatrix}}\in SL(2,\mathbf {C} ),$         (T5)
for Λ = p(λ) ∈ SO(3; 1)$^+$ a Lorentz transformation.
Define
    $[\Pi (\Lambda )P](z)=P[u(\Lambda )(z)],$         (T6)
where P is a Riemann P-function. The resulting function is again a Riemann P-function. The effect of the Mobius transformation of the argument is that of shifting the poles to new locations, hence changing the critical points, but there is no change in the exponents of the differential equation the new function satisfies. The new function is expressed as
    $[\Pi (\Lambda )P](u)=P\left\{{\begin{matrix}\eta &\zeta &\theta &\;\\\alpha &\beta &\gamma &u\\\alpha '&\beta '&\gamma '&\;\end{matrix}}\right\},$         (T6)
where
    $\eta ={\frac {Aa+B}{Ca+D}}\quad {\text{ and }}\quad \zeta ={\frac {Ab+B}{Cb+D}}\quad {\text{ and }}\quad \theta ={\frac {Ac+B}{Cc+D}}.$         (T7)

## Infinite-dimensional unitary representations
### History

The Lorentz group SO(3; 1)$^+$ and its double cover SL(2, C) also have infinite dimensional unitary representations, studied independently by Bargmann (1947), Gelfand & Naimark (1947) and Harish-Chandra (1947) at the instigation of Paul Dirac. This trail of development begun with Dirac (1936) where he devised matrices U and B necessary for description of higher spin (compare Dirac matrices), elaborated upon by Fierz (1939), see also Fierz & Pauli (1939), and proposed precursors of the Bargmann-Wigner equations. In Dirac (1945) he proposed a concrete infinite-dimensional representation space whose elements were called expansors as a generalization of tensors. These ideas were incorporated by Harish–Chandra and expanded with expinors as an infinite-dimensional generalization of spinors in his 1947 paper.
The Plancherel formula for these groups was first obtained by Gelfand and Naimark through involved calculations. The treatment was subsequently considerably simplified by Harish-Chandra (1951) and Gelfand & Graev (1953), based on an analogue for SL(2, C) of the integration formula of Hermann Weyl for compact Lie groups. Elementary accounts of this approach can be found in Rühl (1970) and Knapp (2001).
The theory of spherical functions for the Lorentz group, required for harmonic analysis on the 3-dimensional unit quasi-sphere in Minkowski space, or equivalently 3-dimensional hyperbolic space, is considerably easier than the general theory. It only involves representations from the spherical principal series and can be treated directly, because in radial coordinates the Laplacian on the hyperboloid is equivalent to the Laplacian on R. This theory is discussed in Takahashi (1963), Helgason (1968), Helgason (2000) and the posthumous text of Jorgenson & Lang (2008).

### Principal series for SL(2, C)

The principal series, or unitary principal series, are the unitary representations induced from the one-dimensional representations of the lower triangular subgroup B of G = SL(2, C). Since the one-dimensional representations of B correspond to the representations of the diagonal matrices, with non-zero complex entries z and z$^{−1}$, they thus have the form
    $\chi _{\nu ,k}{\begin{pmatrix}z&0\\c&z^{-1}\end{pmatrix}}=r^{i\nu }e^{ik\theta },$
for k an integer, ν real and with z = re$^{iθ}$. The representations are irreducible; the only repetitions, i.e. isomorphisms of representations, occur when k is replaced by −k. By definition the representations are realized on L$^2$ sections of line bundles on G/B = S$^2$, which is isomorphic to the Riemann sphere. When k = 0, these representations constitute the so-called spherical principal series.
The restriction of a principal series to the maximal compact subgroup K = SU(2) of G can also be realized as an induced representation of K using the identification G / B = K / T, where T = B ∩ K is the maximal torus in K consisting of diagonal matrices with | z | = 1. It is the representation induced from the 1-dimensional representation z$^k$ T, and is independent of ν. By Frobenius reciprocity, on K they decompose as a direct sum of the irreducible representations of K with dimensions |k| + 2m + 1 with m a non-negative integer.
Using the identification between the Riemann sphere minus a point and C, the principal series can be defined directly on L$^2$(C) by the formula
Template:Equation box 1
Irreducibility can be checked in a variety of ways:
* The representation is already irreducible on B. This can be seen directly, but is also a special case of general results on irreducibility of induced representations due to François Bruhat and George Mackey, relying on the Bruhat decomposition G = B ∪ B s B where s is the Weyl group element
    ${\begin{pmatrix}0&-1\\1&0\end{pmatrix}}$.
* The action of the Lie algebra ${\mathfrak {g}}$ of G can be computed on the algebraic direct sum of the irreducible subspaces of K can be computed explicitly and the it can be verified directly that the lowest-dimensional subspace generates this direct sum as a ${\mathfrak {g}}$-module.

### Complementary series for SL(2, C)

The for 0 < t < 2, the complementary series is defined on L$^2$ functions f on C for the inner product
    $(f,g)=\int \int {f(z){\overline {g(w)}}\,dz\,dw \over |z-w|^{2-t}}.$
with the action given by
Template:Equation box 1
The representations in the complementary series are irreducible and pairwise non-isomorphic. As a representation of K, each is isomorphic to the Hilbert space direct sum of all the odd dimensional irreducible representations of K = SU(2). Irreducibility can be proved by analyzing the action of ${\mathfrak {g}}$ on the algebraic sum of these subspaces or directly without using the Lie algebra.

### Plancherel theorem for SL(2, C)

The only irreducible unitary representations of SL(2, C) are the principal series, the complementary series and the trivial representation. Since −I acts as (−1)$^k$ on the principal series and trivially on the remainder, these will give all the irreducible unitary representations of the Lorentz group, provided k is taken to be even.
To decompose the left regular representation of G on L$^2$(G), only the principal series are required. This immediately yields the decomposition on the subrepresentations L$^2$(G/±I), the left regular representation of the Lorentz group, and L$^2$(G/K), the regular representation on 3-dimensional hyperbolic space. (The former only involves principal series representations with k even and the latter only those with k = 0.)
The left and right regular representation λ and ρ are defined on L$^2$(G) by
    $\lambda (g)f(x)=f(g^{-1}x),\,\,\rho (g)f(x)=f(xg).$
Now if f is an element of C$_c$(G), the operator π$_{ν,k}$(f) defined by
    $\pi _{\nu ,k}(f)=\int _{G}f(g)\pi (g)\,dg$
is Hilbert–Schmidt. Define a Hilbert space H by
    $H=\bigoplus _{k\geq 0}HS(L^{2}(C))\otimes L^{2}(R,c_{k}(\nu ^{2}+k^{2})^{1/2}d\nu ),$
where
    $c_{0}=1/4\pi ^{3/2},\,\,c_{k}=1/(2\pi )^{3/2}\,\,(k\neq 0)$
and HS(L$^2$(C)) denotes the Hilbert space of Hilbert–Schmidt operators on L$^2$(C). Then the map U defined on C$_c$(G) by
    $U(f)(\nu ,k)=\pi _{\nu ,k}(f)$
extends to a unitary of L$^2$(G) onto H.
The map U satisfies the intertwining property
    $U(\lambda (x)\rho (y)f)(\nu ,k)=\pi _{\nu ,k}(x)^{-1}\pi _{\nu ,k}(f)\pi _{\nu ,k}(y).$
If f$_1$, f$_2$ are in C$_c$(G) then by unitarity
Template:Equation box 1
Thus if f = f$_1$ ∗ f$_2$* denotes the convolution of f$_1$ and f$_2$*, and $f_{2}^{*}(g)={\overline {f_{2}(g^{-1})}}$, then
Template:Equation box 1
The last two displayed formulas are usually referred to as the Plancherel formula and the Fourier inversion formula respectively. The Plancherel formula extends to all f$_i$ in L$_2$(G). By a theorem of Jacques Dixmier and Paul Malliavin, every function f in $C_{c}^{\infty }(G)$ is a finite sum of convolutions of similar functions, the inversion formula holds for such f. It can be extended to much wider classes of functions satisfying mild differentiability conditions.

### Classification of representations of SO(3, 1)

The strategy followed in the classification of the irreducible infinite-dimensional representations is, in analogy to the finite-dimensional case, to assume they exist, and to investigate their properties. Thus first assume that an irreducible strongly continuous infinite-dimensional representation Π$_H$ on a Hilbert space H of SO(3; 1)$^+$ is at hand. Since SO(3) is a subgroup, Π$_H$ is a representation of it as well. Each irreducible subrepresentation of SO(3) is finite-dimensional, and the SO(3) representation is reducible into a direct sum of irreducible finite-dimensional unitary representations of SO(3) if Π$_H$ is unitary.
The steps are the following:
* Chose a suitable basis of common eigenvectors of J$^2$ and J$_3$.
* Compute matrix elements of J$_1$, J$_2$, J$_3$ and K$_1$, K$_2$, K$_3$.
* Enforce Lie algebra commutation relations.
* Require unitarity together with orthonormality of the basis.

#### Step 1

One may suitably choose a basis and label the basis vectors by
    $\left|j_{0}\,j_{1};j\,m\right\rangle .$
If this was a finite-dimensional representation, then j$_0$ would correspond the lowest occurring eigenvalue j(j + 1) of J$^2$ in the representation, equal to |m − n|, and j$_1$ would correspond to the highest occurring eigenvalue, equal to m + n. In the infinite-dimensional case, j$_0$ ≥ 0 retains this meaning, but j$_1$ does not. One assumes for simplicity that a given j occurs at most once in a given representation (this is the case for finite-dimensional representations), and it can be shown that the assumption is possible to avoid (with a slightly more complicated calculation) with the same results.

#### Step 2

The next step is to compute the matrix elements of the operators J$_1$, J$_2$, J$_3$ and K$_1$, K$_2$, K$_3$ forming the basis of the Lie algebra of so(3; 1). The matrix elements of
    $J_{\pm }=J_{1}\pm iJ_{2},J_{3}$
(here one is operating in the comlpexified Lie algebra) are known from the representation theory of the rotation group, and are given by
    $\left\langle j\,m\right|J_{+}\left|j\,m-1\right\rangle =\left\langle j\,m-1\right|J_{-}\left|j\,m\right\rangle ={\sqrt {(j+m)(j-m+1)}},\quad \left\langle j,m\right|J_{3}\left|j\,m\right\rangle =m,$
where the labels j$_0$ and j$_1$ have been dropped since they are the same for all basis vectors in the representation.
Due to the commutation relations
    $[J_{i},K_{j}]=i\epsilon _{ijk}K_{k},$
the triple (K$_i$, K$_i$, K$_i$) ≡ K is a vector operator and the Wigner–Eckart theorem applies for computation of matrix elements between the states represented by the chosen basis. The matrix elements of
    ${\begin{aligned}K_{0}^{(1)}&=K_{3}\\K_{\pm 1}^{(1)}&=\mp {\frac {1}{\sqrt {2}}}(K_{1}\pm iK_{2}),\end{aligned}}$
where the superscript (1) signifies that the defined quantities are the components of a spherical tensor operator of rank k = 1 (which explains the factor √2 as well) and the subscripts 0, ±1 are referred to as q in formulas below, are given by
    ${\begin{aligned}\left\langle j'\,m'\right|K_{0}^{(1)}\left|j\,m\right\rangle &=\langle j'\,m'\,k=1\,q=0|j\,m\rangle \langle j\|K^{(1)}\|j'\rangle \\\left\langle j'm'\right|K_{\pm 1}^{(1)}\left|j\,m\right\rangle &=\langle j'\,m'\,k=1\,q=\pm 1|j\,m\rangle \langle j\|K^{(1)}\|j'\rangle .\end{aligned}}$
Here the first factors on the right hand sides are Clebsch–Gordan coefficients for coupling j′ with k to get j. The second factors are the reduced matrix elements. They do not depend on m, m′ or q, but depend on j, j′ and, of course, K. For a complete list of non-vanishing equations, see Harish-Chandra (1947, p. 375).

#### Step 3

The next step is to demand that the Lie algebra relations hold, i.e. that
    $[K_{\pm },K_{3}]=\pm J_{\pm },\quad [K_{+},K_{-}]=-2J_{3}.$
This results in a set of equations for which the solutions are
    ${\begin{aligned}\langle j\|K^{(1)}\|j\rangle =i{\frac {j_{1}j_{0}}{\sqrt {j(j+1)}}},\\\langle j\|K^{(1)}\|j-1\rangle =-B_{j}\xi _{j}{\sqrt {j(2j-1)}},\\\langle j-1\|K^{(1)}\|j\rangle =B_{j}\xi _{j}^{-1}{\sqrt {j(2j+1)}},\\\end{aligned}}$
where
    $B_{j}={\sqrt {\frac {(j^{2}-j_{0}^{2})(j^{2}-j_{1}^{2})}{j^{2}(4j^{2}-1)}}},$
and
    ${\begin{aligned}j_{0}&=0,{\frac {1}{2}},1,\ldots ,\\j_{1}&\in \mathbf {C} ,\\\xi _{j}&\in \mathbf {C} .\\\end{aligned}}$

#### Step 4

The imposition of the requirement of unitarity of the corresponding representation of the group restricts the possible values for the arbitrary complex numbers j$_0$ and ξ$_j$. Unitarity of the group representation translates to the requirement of the Lie algebra representatives being Hermitian, meaning
    ${\begin{aligned}K_{\pm }^{\dagger }&=K_{\mp },\\K_{3}^{\dagger }&=K_{3}.\end{aligned}}$
This translates to
    ${\begin{aligned}\langle j\|K^{(1)}\|j\rangle &={\overline {\langle j\|K^{(1)}\|j\rangle }},\\\langle j\|K^{(1)}\|j-1\rangle &=-{\overline {\langle j-1\|K^{(1)}\|j\rangle }},\\\end{aligned}}$
leading to
    ${\begin{aligned}j_{0}(j_{1}+{\overline {j_{1}}})&=0,\\|B_{j}|(|\xi _{j}|^{2}-e^{-2i\beta _{j}})&=0,\end{aligned}}$
where β$_j$ is the angle of B$_j$ on polar form. For |B$_j$| ≠ 0 one has |ξ|Template:Supsub = 1, and ξ$_j$ = 1 is chosen by convention. There are two possible cases. The first with j$_1$ + j$_1$ = 0 gives, with j$_1$ = − iν, ν real,
    ${\begin{aligned}\langle j\|K^{(1)}\|j\rangle &={\frac {\nu j_{0}}{j(j+1)}},\\B_{j}&={\sqrt {\frac {(j^{2}-j_{0}^{2})(j^{2}+\nu ^{2})}{4j^{2}-1}}}.\end{aligned}}$
This is principal series and the elements may be denoted (j$_0$, ν), 2j$_0$ ∈ ℕ, ν ∈ R. For the other possibility, j$_0$ = 0, one has
    ${\begin{aligned}\langle j\|K^{(1)}\|j\rangle &=0,\\B_{j}&={\sqrt {\frac {(j^{2}-\nu ^{2})}{4j^{2}-1}}}.\end{aligned}}$
One needs to require that BTemplate:Supsub is real and positive for j = 1, 2, ... (because B$_0$ = B$_{j0}$), leading to −1 ≤ ν ≤ 1. This is complementary series and its elements may be denoted (0, ν), −1 ≤ ν ≤ 1.
This shows that the representations of above are all infinite-dimensional irreducible unitary representations.

## Explicit formulas
### Conventions and Lie algebra bases

The metric of choice is given by η = diag(−1, 1, 1, 1), and the physics convention for Lie algebras and the exponential mapping is used. These choices are arbitrary, but once they are made, fixed. One possible choice of basis for the Lie algebra is, in the 4-vector representation, given by
    ${\begin{aligned}J_{1}&=J^{23}=-J^{32}=i{\biggl{\begin{smallmatrix}0&0&0&0\\0&0&0&0\\0&0&0&-1\\0&0&1&0\\\end{smallmatrix}}{\biggr )},\\J_{2}&=J^{31}=-J^{13}=i{\biggl{\begin{smallmatrix}0&0&0&0\\0&0&0&1\\0&0&0&0\\0&-1&0&0\\\end{smallmatrix}}{\biggr )},\\J_{3}&=J^{12}=-J^{21}=i{\biggl{\begin{smallmatrix}0&0&0&0\\0&0&-1&0\\0&1&0&0\\0&0&0&0\\\end{smallmatrix}}{\biggr )},\\K_{1}&=J^{01}=J^{10}=i{\biggl{\begin{smallmatrix}0&1&0&0\\1&0&0&0\\0&0&0&0\\0&0&0&0\\\end{smallmatrix}}{\biggr )},\\K_{2}&=J^{02}=J^{20}=i{\biggl{\begin{smallmatrix}0&0&1&0\\0&0&0&0\\1&0&0&0\\0&0&0&0\\\end{smallmatrix}}{\biggr )},\\K_{3}&=J^{03}=J^{30}=i{\biggl{\begin{smallmatrix}0&0&0&1\\0&0&0&0\\0&0&0&0\\1&0&0&0\\\end{smallmatrix}}{\biggr )}.\end{aligned}}$
The commutation relations of the Lie algebra so(3; 1) are
    $[J^{\mu \nu },J^{\rho \sigma }]=i(\eta ^{\sigma \mu }J^{\rho \nu }+\eta ^{\nu \sigma }J^{\mu \rho }-\eta ^{\rho \mu }J^{\sigma \nu }-\eta ^{\nu \rho }J^{\mu \sigma }).$
In three-dimensional notation, these are
    $[J_{i},J_{j}]=i\epsilon _{ijk}J_{k},\quad [J_{i},K_{j}]=i\epsilon _{ijk}K_{k},\quad [K_{i},K_{j}]=-i\epsilon _{ijk}J_{k}.$
The choice of basis above satisfies the relations, but other choices are possible. The multiple use of the symbol J above and in the sequel should be observed.

### Weyl spinors and bispinors

By taking, in turn, m = 1/2, n = 0 and m = 0, n = 1/2 and by setting
    $J_{i}^{({\frac {1}{2}})}={\frac {1}{2}}\sigma _{i}$
in the general expression (G1), and by using the trivial relations 1$_1$ = 1 and J$^{(0)}$ = 0, one obtains
    ${\begin{aligned}\pi _{({\frac {1}{2}},0)}(J_{i})&={\frac {1}{2}}(\sigma _{i}\otimes 1_{(1)}+1_{(2)}\otimes J_{i}^{(0)})={\frac {1}{2}}\sigma _{i}\quad \pi _{({\frac {1}{2}},0)}(K_{i})={\frac {i}{2}}(1_{(2)}\otimes J_{i}^{(0)}-\sigma _{i}\otimes 1_{(1)})=-{\frac {i}{2}}\sigma _{i},\\\pi _{(0,{\frac {1}{2}})}(J_{i})&={\frac {1}{2}}(J_{i}^{(0)}\otimes 1_{(2)}+1_{(1)}\otimes \sigma _{i})={\frac {1}{2}}\sigma _{i}\quad \pi _{(0,{\frac {1}{2}})}(K_{i})={\frac {i}{2}}(1_{(1)}\otimes \sigma _{i}-J_{i}^{(0)}\otimes 1_{(2)})=+{\frac {i}{2}}\sigma _{i}.\end{aligned}}$         (W1)
These are the left-handed and right-handed Weyl spinor representations. They act by matrix multiplication on 2-dimensional complex vector spaces (with a choice of basis) V$_L$ and V$_R$, whose elements Ψ$_L$ and Ψ$_R$ are called left- and right-handed Weyl spinors respectively. Given (π(1/2,0), V$_L$) and (π(0,1/2), V$_R$) one may form their direct sum as representations,
    ${\begin{aligned}\pi _{({\frac {1}{2}},0)\oplus (0,{\frac {1}{2}})}(J_{i})&={\frac {1}{2}}{\biggl{\begin{matrix}\sigma _{i}&0\\0&\sigma _{i}\\\end{matrix}}{\biggr )},\\\pi _{({\frac {1}{2}},0)\oplus (0,{\frac {1}{2}})}(K_{i})&={\frac {i}{2}}{\biggl{\begin{matrix}\sigma _{i}&0\\0&-\sigma _{i}\\\end{matrix}}{\biggr )}\\\end{aligned}}.$         (D1)
This is, up to a similarity transformation, the (1/2,0) ⊕ (0,1/2) Dirac spinor representation of so(3; 1). It acts on the 4-component elements (Ψ$_L$, Ψ$_R$) of (V$_L$ ⊕ V$_R$), called bispinors, by matrix multiplication. The representation may be obtained in a more general and basis independent way using Clifford algebras. These expressions for bispinors and Weyl spinors all extend by linearity of Lie algebras and representations to all of so(3; 1). Expressions for the group representations are obtained by exponentiation.